% This file contains a LaTeX version of the joint % PYTHIA 5.7 and JETSET 7.4 % physics description and manual % Copyright Torbjorn Sjostrand 1993 \documentstyle[12pt]{article} %define page size \setlength{\textheight}{245mm} \setlength{\topmargin}{-5mm} \setlength{\headheight}{0mm} \setlength{\headsep}{0mm} \setlength{\footheight}{10mm} \setlength{\textwidth}{160mm} \setlength{\oddsidemargin}{0mm} \setlength{\evensidemargin}{0mm} %line spacing \renewcommand{\baselinestretch}{0.9} %define math alphabets for roman, boldface and teletype \newcommand{\mrm}[1]{\mathrm{#1}} \newmathalphabet*{\mbf}{cmr}{b}{n} \newmathalphabet*{\mtt}{cmtt}{m}{n} %program names \newcommand{\Py}{\mbox{\sc Pythia}} \newcommand{\Je}{\mbox{\sc Jetset}} \newcommand{\PyJe}{\mbox{\sc Pythia/Jetset}} \newcommand{\JePy}{\mbox{\sc Jetset/Pythia}} %some frequent physics symbols \newcommand{\alphas}{\alpha_{\mrm{s}}} \newcommand{\alphaem}{\alpha_{\mrm{em}}} \newcommand{\ssintw}{\sin^2 \! \theta_W} 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\newcommand{\itemn}[1]{\item[{#1}\hfill]} %draw box around routine or common block name \newcommand{\drawbox}[1]{\vspace{\baselineskip}\noindent% \fbox{\hspace{3mm}\rule[-2mm]{0mm}{6.5mm}{\tt #1}\hspace{3mm}}% \vspace{0.5\baselineskip}} \newcommand{\drawboxtwo}[2]{\vspace{\baselineskip}\noindent% \fbox{\hspace{3mm}\rule{0mm}{7.0mm}\begin{minipage}{150mm}% \begin{tabbing}{\tt #1}\\{\tt #2}\end{tabbing}\end{minipage}% \rule[-4.5mm]{0mm}{4.5mm}\hspace{3mm}}\vspace{0.5\baselineskip}} \newcommand{\drawboxthree}[3]{\vspace{\baselineskip}\noindent% \fbox{\hspace{3mm}\rule{0mm}{9.7mm}\begin{minipage}{150mm}% \begin{tabbing}{\tt #1}\\{\tt #2}\\{\tt #3}\end{tabbing}% \end{minipage}\rule[-7.2mm]{0mm}{7.2mm}\hspace{3mm}}% \vspace{0.5\baselineskip}} \newcommand{\boxsep}{\vspace{0.5\baselineskip}} %fraction of page devoted to figures \setcounter{topnumber}{1} \setcounter{bottomnumber}{1} \renewcommand{\textfraction}{0.1} \renewcommand{\topfraction}{0.9} \renewcommand{\bottomfraction}{0.9} \renewcommand{\floatpagefraction}{0.8} %one caption and figure/table, indented but spanning whole page \newlength{\captivewidth} \setlength{\captivewidth}{\textwidth} \addtolength{\captivewidth}{-10mm} \newcommand{\captive}[1]{\rule{5mm}{0mm}% \begin{minipage}{\captivewidth}% \caption[small]{#1}\end{minipage}} \newcommand{\onefigure}[2]{\begin{figure}[tb]\vspace{#1} \captive{#2}\end{figure}} %minimal separation between rows in table \newlength{\tablinsep} \setlength{\tablinsep}{0.73\baselineskip} %size of column (half page width + blank in middle) \newlength{\halfpagewid} \setlength{\halfpagewid}{0.5\textwidth} \addtolength{\halfpagewid}{-10mm} %set level for table of contents \setcounter{tocdepth}{2} \begin{document} %set sloppy attitude to line breaks \sloppy \pagestyle{empty} \begin{flushright} CERN-TH.7112/93 \\[0.5\baselineskip] \fbox{W5035/W5044} \end{flushright} \vspace{\fill} \begin{center} {\Huge\bf P}{\LARGE\bf YTHIA} \ {\Huge\bf 5.7} \ {\LARGE\bf and} \ % {\Huge\bf J}{\LARGE\bf ETSET} \ {\Huge\bf 7.4} \\[5mm] {\LARGE\bf Physics and Manual} \\[10mm] {\Large\bf Torbj\"orn Sj\"ostrand} \\[3mm] {\large Theory Division, CERN} \\[1mm] {\large CH-1211 Geneva 23}\\[1mm] {\large Switzerland}\\ \vspace{\fill} \fbox{\hspace{3mm}\rule{0mm}{8mm} \begin{minipage}[t]{150mm} \begin{tabbing} {\tt ~~~~~~~~~~~~*......*~~~~~~~~~~~~}\\[-1.1mm] {\tt ~~~~~~~*:::!!:::::::::::*~~~~~~~}\\[-1.1mm] {\tt ~~~~*::::::!!::::::::::::::*~~~~}\\[-1.1mm] {\tt ~~*::::::::!!::::::::::::::::*~~}\\[-1.1mm] {\tt ~*:::::::::!!:::::::::::::::::*~}\\[-1.1mm] {\tt ~*:::::::::!!:::::::::::::::::*~}\\[-1.1mm] {\tt ~~*::::::::!!::::::::::::::::*!~}\\[-1.1mm] {\tt ~~~~*::::::!!::::::::::::::*~!!~}\\[-1.1mm] {\tt ~~~~!!~*:::!!:::::::::::*~~~~!!~}\\[-1.1mm] {\tt ~~~~!!~~~~~!*~-><-~*~~~~~~~~~!!~}\\[-1.1mm] {\tt ~~~~!!~~~~~!!~~~~~~~~~~~~~~~~!!~}\\[-1.1mm] {\tt ~~~~!!~~~~~!!~~~~~~~~~~~~~~~~!!~}\\[-1.1mm] {\tt ~~~~!!~~~~~~~~~~~~~~~~~~~~~~~!!~}\\[-1.1mm] {\tt ~~~~!!~~~~~~~~ep~~~~~~~~~~~~~!!~}\\[-1.1mm] {\tt ~~~~!!~~~~~~~~~~~~~~~~~~~~~~~!!~}\\[-1.1mm] {\tt ~~~~!!~~~~~~~~~~~~~~~~~pp~~~~!!~}\\[-1.1mm] {\tt ~~~~!!~~~e+e-~~~~~~~~~~~~~~~~!!~}\\[-1.1mm] {\tt ~~~~!!~~~~~~~~~~~~~~~~~~~~~~~!!~}\\[-1.1mm] {\tt ~~~~!!~~~~~~~~~~~~~~~~~~~~~~~~~~}\\[-1.1mm] \end{tabbing} \end{minipage} \rule[-6mm]{0mm}{6mm}\hspace{3mm}} \end{center} \vspace*{\fill} \noindent CERN-TH.7112/93 \\ December 1993 \clearpage \vspace*{\fill} \noindent {\hfill\framebox[0.8\textwidth][t]{ \begin{minipage}{0.75\textwidth} \vspace{\baselineskip} \begin{center} Copyright Notice \end{center} \vspace{\baselineskip} {\bf CERNLIB -- CERN Program Library Long writeups} \vspace{\baselineskip} \copyright{} Copyright CERN, Geneva 1993 \vspace{\baselineskip} Copyright and any other appropriate legal protection of these computer programs and associated documentation reserved in all countries of the world. \vspace{\baselineskip} These programs or documentation may not be reproduced by any method without prior written consent of the Director-General of CERN or his delegate. \vspace{\baselineskip} Permission for the usage of any programs described herein is granted apriori to those scientific institutes associated with the CERN experimental program or with whom CERN has concluded a scientific collaboration agreement. \vspace{\baselineskip} Requests for information should be addressed to: \vspace{-0.5\baselineskip} \begin{center} \begin{tabular}{l@{\protect\rule{0mm}{\tablinsep}}} {\tt CERN Program Library Office } \\ {\tt CERN-CN Division } \\ {\tt CH-1211 Geneva 23 } \\ {\tt Switzerland } \\ {\tt Tel. +41 22 767 4951 } \\ {\tt Fax. +41 22 767 7155 } \\ {\tt Bitnet: CERNLIB@CERNVM } \\ {\tt DECnet: VXCERN::CERNLIB (node 22.190) } \\ {\tt Internet: CERNLIB@CERNVM.CERN.CH } \end{tabular} \end{center} \vspace{0mm} \end{minipage}}\hfill} \vspace{2\baselineskip}\noindent {\hfill\begin{minipage}{0.8\textwidth} {\bf Trademark notice: All trademarks appearing in this guide are acknowledged as such.} \end{minipage}\hfill} \vspace*{\fill} \cleardoublepage \section*{Preface} The {\Py} and {\Je} programs are frequently used for event generation in high-energy physics. The emphasis is on multiparticle production in collisions between elementary particles. This in particular means hard interactions in $\ee$, $\pp$ and $\ep$ colliders, although also other applications are envisaged. The programs are intended to generate complete events, in as much detail as experimentally observable ones, within the bounds of our current understanding of the underlying physics. Many of the components of the programs represent original research, in the sense that models have been developed and implemented for a number of aspects not covered by standard theory. Although originally conceived separately, the {\Py} and {\Je} programs today are so often used together that it makes sense to present them here without too much distinction. Both programs have a long history, and several manuals have come out. The previous round of {\PyJe} program descriptions is by now over five years old. Meanwhile a large number of additions and changes have been made. The manual aspect has been covered by companion computer files to the programs, while the physics documentation has lagged behind. In the current report I try to provide an up-to-date description of both these aspects of the two programs. Currently this report is only preliminary, however. Until further notice, you are therefore asked to quote \\[1mm] T. Sj\"ostrand, Computer Physics Commun. {\bf 39} (1986) 347; \\ T. Sj\"ostrand and M. Bengtsson, Computer Physics Commun. {\bf 43} (1987) 367 \\[1mm] when you use {\Je}, and \\[1mm] H.-U. Bengtsson and T. Sj\"ostrand, Computer Physics Commun. {\bf 46} (1987) 43 \\[1mm] when you use {\Py}, as before. Also remember to cite the original literature on the physics topics of particular relevance for your studies. (There is no reason to omit references to good physics papers simply because some of their contents have also been made available as program code.) Event generators often have a reputation for being `black boxes'; if nothing else, this report should provide you with a glimpse of what goes on inside the programs. Some such understanding may be of special interest for new users, who have no background in the field. An attempt has been made to structure the report sufficiently well that many of the sections can be read independently of each other, so you can pick the sections that interest you. I have tried to keep together the physics and the manual sections on specific topics, where practicable, which represents a change of policy compared with previous manual versions. Any feedback on this and other aspects is welcome. A large number of persons should be thanked for their contributions. Hans-Uno Bengtsson is the originator of the {\Py} program, and for many years we worked in parallel on its further development. Mats Bengtsson is the main author of the final-state parton-shower algorithm. Bo Andersson and G\"osta Gustafson are the originators of the Lund model, and strongly influenced the early development of the programs. Further comments on the programs have been obtained from users too numerous to be mentioned here, but who are all gratefully acknowledged. To write programs of this size and complexity would be impossible without a strong user feedback. The moral responsibility for any remaining errors clearly rests with me. However, kindly note that this is a `University World' product, distributed `as is', free of charge, without any binding guarantees. And always remember that the programs do not represent a dead collection of established truths, but rather one of many possible approaches to the problem of multiparticle production in high-energy physics, at the frontline of current research. Be critical! \newpage %this page is empty to keep page count right \mbox{} \cleardoublepage \tableofcontents \cleardoublepage \pagestyle{plain} \setcounter{page}{1} \section{Introduction} Multiparticle production is the most characteristic feature of current high-energy physics. Today, observed particle multiplicities are typically between ten and a hundred, and with future machines this range will be extended upwards. The bulk of the multiplicity is found in jets, i.e. in bunches of hadrons (or decay products of hadrons) produced by the hadronization of quarks and gluons. \subsubsection*{The Complexity of High-Energy Processes} To first approximation, all processes have a simple structure at the level of interactions between the fundamental objects of nature, i.e. quarks, leptons and gauge bosons. For instance, a lot can be understood about the structure of hadronic events at LEP just from the `skeleton' process $\ee \to \Z^0 \to \q \qbar$. Corrections to this picture can be subdivided, arbitrarily but conveniently, into three main classes. Firstly, there are bremsstrahlung-type modifications, i.e. the emission of additional final-state particles by branchings such as $\e \to \e \gamma$ or $\q \to \q \g$. Because of the largeness of the strong coupling constant $\alphas$, and because of the presence of the triple gluon vertex, QCD emission off quarks and gluons is especially prolific. We therefore speak about `parton showers', wherein a single initial parton may give rise to a whole bunch of partons in the final state. Also photon emission may give sizeable effects in $\ee$ and $\ep$ processes. The bulk of the bremsstrahlung corrections are universal, i.e. do not depend on the details of the process studied, but only on one or a few key numbers, such as the momentum transfer scale of the process. Such universal corrections may be included to arbitrarily high orders, using a probabilistic language. Alternatively, exact calculations of bremsstrahlung corrections may be carried out order by order in perturbation theory, but rapidly the calculations then become prohibitively complicated and the answers correspondingly lengthy. Secondly, we have `true' higher-order corrections, which involve a combination of loop graphs and the soft parts of the bremsstrahlung graphs above, a combination needed to cancel some divergences. In a complete description it is therefore not possible to consider bremsstrahlung separately, as assumed here. The necessary perturbative calculations are usually very difficult; only rarely have results been presented that include more than one non-`trivial' order, i.e. more than one loop. As above, answers are usually very lengthy, but some results are sufficiently simple to be generally known and used, such as the running of $\alphas$, or the correction factor $1 + \alphas/\pi + \cdots$ in the partial widths of $\Z^0 \to \q \qbar$ decay channels. For high-precision studies it is imperative to take into account the results of loop calculations, but usually effects are minor for the qualitative aspects of high-energy processes. Thirdly, quarks and gluons are confined. In the two points above, we have used a perturbative language to describe the short-distance interactions of quarks, leptons and gauge bosons. For leptons and colourless bosons this language is sufficient. However, for quarks and gluons it must be complemented with a picture for the hadronization process (which can be subdivided into fragmentation and decays), wherein the coloured partons are transformed into jets of colourless hadrons, photons and leptons. This process is still not yet understood from first principles, but has to be based on models. In one sense, hadronization effects are overwhelmingly large, since this is where the bulk of the multiplicity comes from. In another sense, the overall energy flow of a high-energy event is mainly determined by the perturbative processes, with only a minor additional smearing caused by the hadronization step. One may therefore pick different levels of ambition, but in general detailed studies require a detailed modelling of the hadronization process. The simple structure that we started out with has now become considerably more complex --- instead of maybe two final-state partons we have a hundred final particles. The original physics is not gone, but the skeleton process has been dressed up and is no longer directly visible. A direct comparison between theory and experiment is therefore complicated at best, and impossible at worst. \subsubsection*{Event Generators} It is here that event generators come to the rescue. In an event generator, the objective strived for is to use computers to generate events as detailed as could be observed by a perfect detector. This is not done in one step, but rather by `factorizing' the full problem into a number of components, each of which can be handled reasonably accurately. Basically, this means that the hard process is used as input to generate bremsstrahlung corrections, and that the result of this exercise is thereafter left to hadronize. This sounds a bit easier than it really is --- else this report would be a lot thinner. However, the basic idea is there: if the full problem is too complicated to be solved in one go, try to subdivide it into smaller tasks of manageable proportions. In the actual generation procedure, most steps therefore involve the branching of one object into two, or at least into a very small number, each of which being free to branch in its turn. A lot of bookkeeping is involved, but much is of a repetitive nature, and can therefore be left for the computer to handle. As the name indicates, the output of an event generator should be in the form of `events', with the same average behaviour and the same fluctuations as real data. In the data, fluctuations arise from the quantum mechanics of the underlying theory. In generators, Monte Carlo techniques are used to select all relevant variables according to the desired probability distributions, and thereby ensure randomness in the final events. Clearly some loss of information is entailed: quantum mechanics is based on amplitudes, not probabilities. However, only very rarely do (known) interference phenomena appear that cannot be cast in a probabilistic language. This is therefore not a more restraining approximation than many others. Once there, an event generator can be used in many different ways. The five main applications are probably the following: \begin{Itemize} \item To give physicists a feeling for the kind of events one may expect/hope to find, and at what rates. \item As a help in the planning of a new detector, so that detector performance is optimized, within other constraints, for the study of interesting physics scenarios. \item As a tool for devising the analysis strategies that should be used on real data, so that signal-to-background conditions are optimized. \item As a method for estimating detector acceptance corrections that have to be applied to raw data, in order to extract the `true' physics signal. \item As a convenient framework within which to interpret the observed phenomena in terms of a more fundamental underlying theory (usually the Standard Model). \end{Itemize} Where does a generator fit into the overall analysis chain of an experiment? In `real life', the machine produces interactions. These events are observed by detectors, and the interesting ones are written to tape by the data acquisition system. Afterwards the events may be reconstructed, i.e. the electronics signals (from wire chambers, calorimeters, and all the rest) may be translated into a deduced setup of charged tracks or neutral energy depositions, in the best of worlds with full knowledge of momenta and particle species. Based on this cleaned-up information, one may proceed with the physics analysis. In the Monte Carlo world, the r\^ole of the machine, namely to produce events, is taken by the event generators described in this report. The behaviour of the detectors --- how particles produced by the event generator traverse the detector, spiral in magnetic fields, shower in calorimeters, or sneak out through cracks, etc. --- is simulated in programs such as {\sc Geant} \cite{Bru89}. Traditionally, this latter activity is called event simulation, which is somewhat unfortunate since the same words could equally well be applied to what, here, we call event generation. A more appropriate term is detector simulation. Ideally, the output of this simulation has exactly the same format as the real data recorded by the detector, and can therefore be put through the same event reconstruction and physics analysis chain, except that here we know what the `right answer' should be, and so can see how well we are doing. Since the full chain of detector simulation and event reconstruction is very time-consuming, one often does `quick and dirty' studies in which these steps are skipped entirely, or at least replaced by very simplified procedures which only take into account the geometric acceptance of the detector and other trivial effects. One may then use the output of the event generator directly in the physics studies. There are still many holes in our understanding of the full event structure, despite an impressive amount of work and detailed calculations. To put together a generator therefore involved making a choice on what to include, and how to include it. At best, the spread between generators can be used to give some impression of the uncertainties involved. A multitude of approximations will be discussed in the main part of this report, but already here is should be noted that many major approximations are related to the almost complete neglect of the second point above, i.e. of the non-`trivial' higher-order effects. It can therefore only be hoped that the `trivial' higher order parts give the bulk of the experimental behaviour. By and large, this seems to be the case; for $\ee$ annihilation it even turns out to be a very good approximation. The necessity to make compromises has one major implication: to write a good event generator is an art, not an exact science. It is therefore essential not to blindly trust the results of any single event generator, but always to make several cross-checks. In addition, with computer programs of tens of thousands of lines, the question is not whether bugs exist, but how many there are, and how critical their positions. Further, an event generator cannot be thought of as all-powerful, or able to give intelligent answers to ill-posed questions; sound judgement and some understanding of a generator are necessary prerequisites for successful use. In spite of these limitations, the event generator approach is the most powerful tool at our disposal if we wish to gain a detailed and realistic understanding of physics at current or future high-energy colliders. \subsubsection*{The Origins of the JETSET and PYTHIA Programs} Over the years, many event generators have appeared. Surveys of generators for $\ee$ physics in general and LEP in particular may be found in \cite{Kle89,Sjo89}, for high-energy hadron--hadron ($\pp$) physics in \cite{Ans90,Sjo92}, and for $\ep$ physics in \cite{HER92}. We refer the reader to those for additional details and references. In this particular report, the two closely connected programs {\Je} and {\Py} will be described. {\Je} has its roots in the efforts of the Lund group to understand the hadronization process, starting in the late seventies \cite{And83}. The so-called string fragmentation model was developed as an explicit and detailed framework, within which the long-range confinement forces are allowed to distribute the energies and flavours of a parton configuration among a collection of primary hadrons, which subsequently may decay further. This model, known as the Lund string model, or `Lund' for short, contained a number of specific predictions, which were confirmed by data from PETRA and PEP, whence the model gained a widespread acceptance. The Lund string model is still today the most elaborate and widely used fragmentation model at our disposal. It remains at the heart of the {\JePy} programs. In order to predict the shape of events at PETRA/PEP, and to study the fragmentation process in detail, it was necessary to start out from the partonic configurations that were to fragment. The generation of complete $\ee$ hadronic events was therefore added, originally based on simple $\gamma$ exchange and first-order QCD matrix elements, later extended to full $\gammaZ$ exchange with first-order initial-state QED radiation and second-order QCD matrix elements. A number of utility routines were also provided early on, for everything from event listing to jet finding. By the mid-eighties it was clear that the matrix-element approach had reached the limit of its usefulness, in the sense that it could not fully describe the multijet topologies of the data. (Later on, the use of optimized perturbation theory was to lead to a resurgence of the matrix-element approach, but only for specific applications.) Therefore a parton-shower description was developed \cite{Ben87a} as an alternative to the matrix-element one. The combination of parton showers and string fragmentation has been very successful, and forms the main approach to the description of hadronic $\Z^0$ events. In recent years, {\Je} has been a fairly stable product, covering the four main areas of fragmentation, final-state parton showers, $\ee$ event generation and general utilities. The successes of string fragmentation in $\ee$ made it interesting to try to extend this framework to other processes, and explore possible physics consequences. Therefore a number of other programs were written, which combined a process-specific description of the hard interactions with the general fragmentation framework of {\Je}. The {\Py} program evolved out of early studies on fixed-target proton--proton processes, addressed mainly at issues related to string drawing. With time, the interest shifted towards higher energies, first to the SPS $\ppbar$ collider, and later to SSC and LHC, in the context of a number of workshops in the USA and Europe. Parton showers were added, for final-state radiation by making use of the {\Je} routine, for initial-state one by the development of the concept of `backwards evolution', specifically for {\Py} \cite{Sjo85}. Also a framework was developed for minimum-bias and underlying events \cite{Sjo87a}. Another main change was the introduction of an increasing number of hard processes, within the Standard Model and beyond. A special emphasis was put on the search for the Standard Model Higgs, in different mass ranges and in different channels, with due respect to possible background processes. The bulk of the machinery developed for hard processes actually depended little on the choice of initial state, as long as the appropriate parton distributions were there for the incoming partons and particles. It therefore made sense to extend the program from being only a $\pp$ generator to working also for $\ee$ and $\ep$. This process was only completed in 1991, again spurred on by physics workshop activities. Currently {\Py} should therefore work equally well for a selection of different possible incoming beam particles. The tasks of including new processes, and of improving the simulation of already present ones, are never-ending. Work therefore continues apace. While {\Je} still is formally independent of {\Py}, their ties have grown much stronger over the years, and the border-line between the two programs has become more and more artificial. It is no coincidence that the two are presented together here; this way a lot of repetition of common material can be avoided. The price to be paid is that some differences in philosophy will have to be discussed. \subsubsection*{About this Report} As we see, {\Je} and {\Py} started out as very ideologically motivated programs, developed to study specific physics questions in enough detail that explicit predictions could be made for experimental quantities. As it was recognized that experimental imperfections could distort the basic predictions, the programs were made available for general use by experimentalists. It thus became feasible to explore the models in more detail than would otherwise have been possible. As time went by, the emphasis came to shift somewhat, away from the original strong coupling to a specific fragmentation model, towards a description of high-energy multiparticle production processes in general. Correspondingly, the use expanded from being one of just comparing data with specific model predictions, to one of extensive use for the understanding of detector performance, for the derivation of acceptance correction factors, for the prediction of physics at future high-energy accelerators, and for the design of related detectors. While the ideology may be less apparent, it is still there, however. This is not something unique to the programs discussed here, but inherent in any event generator, or at least any generator that attempts to go beyond the simple parton level skeleton description of a hard process. Do not accept the myth that everything available in Monte Carlo form represents ages-old common knowledge, tested and true. Ideology is present by commissions or omissions in any number of details. Programs like {\Py} and {\Je} represent a major amount of original physics research, often on complicated topics where no simple answers are available. As a (potential) program user you must be aware of this, so that you can form your own opinion, not just about what to trust and what not to trust, but also how much to trust a given prediction, i.e. how uncertain it is likely to be. {\Je} and {\Py} are particularly well endowed in this respect, since a number of publications exist where most of the relevant physics is explained in considerable detail. In fact, the problem may rather be the opposite, to find the relevant information among all the possible places. One main objective of the current report is therefore to collect much of this information in one single place. Not all the material found in specialized papers is reproduced, by a wide margin, but at least enough should be found here to understand the general picture and to know where to go for details. The current report is therefore intended eventually to replace the previous round of published physics descriptions and program manuals. However, we are not there yet. The correct references to {\Je} still are \cite{Sjo86} and \cite{Sjo87} and to {\Py} \cite{Ben87}. Further specification could include a statement of the type `We use {\Py} version X.x and {\Je} version Y.y'. (If you are a {\LaTeX} fan, you may want to know that the program names in this report have been generated by the commands \verb+{\sc Jetset}+ and \verb+{\sc Pythia}+.) Kindly do not refer to {\JePy} as `unpublished', `private communication' or `in preparation': such phrases are only creating unnecessary confusion. In addition, remember that many of the individual physics components are documented in separate publications. If some of these contain ideas that are useful to you, there is every reason to cite them. A reasonable selection would vary as a function of the physics you are studying. The criterion for which to pick should be simple: imagine that a Monte Carlo implementation had not been available. Would you then have cited a given paper on the grounds of its physics contents alone? If so, do not punish the extra effort of turning these ideas into publicly available software. (Monte Carlo manuals are good for nothing in the eyes of many theorists, so often only the acceptance of `mainstream' publications counts.) Here follows a list of some main areas where the $\PyJe$ programs contain original research: \begin{Itemize} \item The string fragmentation model \cite{And83}. \item The string effect \cite{And80}. \item Baryon production (diquark/popcorn) \cite{And82,And85}. \item Fragmentation of multiparton systems \cite{Sjo84}. \item Fragmentation effects on $\alphas$ determinations \cite{Sjo84a}. \item Initial state parton showers \cite{Sjo85}. \item Final state parton showers \cite{Ben87a}. \item Photon radiation from quarks \cite{Sjo92c} \item Deep inelastic scattering \cite{And81a,Ben88}. \item Photoproduction \cite{Sch93a}. \item Colour flow in hard scatterings \cite{Ben84}. \item Elastic and diffractive cross sections \cite{Sch93b}. \item Minijets (multiple parton--parton interactions) \cite{Sjo87a}. \item Rapidity gaps \cite{Dok92}. \item Jet clustering in $k_{\perp}$ \cite{Sjo83}. \end{Itemize} In addition to a physics survey, the current report also contains a complete manual for the two programs. Such manuals have always been updated and distributed jointly with the programs. To a first approximation, we therefore do not have much new to offer here. However, an attempt has been made to group the material more logically according to physics topics than in previous distributions, to tie it closer to the physics description, and to improve the layout and therefore the readability. Any feedback is welcome. A word of warning may be in place. The program description is fairly lengthy, and certainly could not be absorbed in one sitting. This is not even necessary, since all switches and parameters are provided with sensible default values, based on our best understanding (of the physics, and of what you expect to happen if you do not specify any options). As a new user, you can therefore disregard all the fancy options, and just run the program with a minimum ado. Later on, as you gain experience, the options that seem useful can be tried out. No single user is ever likely to find need for more than a fraction of the total number of possibilities available, yet many of them have been added to meet specific user requests. \subsubsection*{Disclaimer} At all times it should be remembered that this is not a commercial product, developed and supported by professionals. Instead it is a `University World' product, developed by a very few physicists (mainly the current author) originally for their own needs, and supplied to other physicists on an `as-is' basis, free of charge. No guarantees are therefore given for the proper functioning of the programs, nor for the validity of physics results. In the end, it is always up to you to decide for yourself whether to trust a given result or not. Usually this requires comparison either with analytical results or with results of other programs, or with both. Even this is not necessarily foolproof: for instance, if an error is made in the calculation of a matrix element for a given process, this error will be propagated both into the analytical results based on the original calculation and into all the event generators which subsequently make use of the published formulae. In the end, there is no substitute for a sound physics judgement. This does not mean that you are all on your own, with a program nobody feels responsible for. Attempts are made to check processes as carefully as possible, to write programs that do not invite unnecessary errors, and to provide a detailed and accurate documentation. All of this while maintaining the full power and flexibility, of course, since the physics must always take precedence in any conflict of interests. If nevertheless any errors or unclarities are found, please do communicate them to me, e.g. on phone +41 - 22 - 767 28 20 or BITNET TORSJO@CERNVM. Every attempt will be made to solve problems as soon as is reasonably possible, given that this support is by one person alone, who also has other responsibilities. \subsubsection*{Appendix: The Historical Pythia} While the origin and connotations of the `{\Je}' program name should be commonly known, the `{\Py}' label may need some explanation. The myth tells how Apollon, the God of Wisdom, killed the powerful dragon-like monster Python, close to the village of Delphi in Greece. To commemorate this victory, Apollon founded the Pythic Oracle in Delphi, on the slopes of Mount Parnassos. Here men could come to learn the will of the Gods and the course of the future. The oracle plays an important r\^ole in many of the other Greek myths, such as those of Heracles and of King Oedipus. Questions were to be put to the Pythia, the `Priestess' or `Prophetess' of the Oracle. In fact, she was a local woman, usually a young maiden, of no particular religious schooling. Seated on a tripod, she inhaled the obnoxious vapours that seeped up through a crevice in the ground. This brought her to a trance-like state, in which she would scream seemingly random words and sounds. It was the task of the professional priests in Delphi to record those utterings and edit them into the official Oracle prophecies, which often took the form of poems in perfect hexameter. In fact, even these edited replies were often less than easy to interpret. The Pythic oracle acquired a reputation for ambiguous answers. The Oracle existed already at the beginning of the historical era in Greece, and was universally recognized as the foremost religious seat. Individuals and city states came to consult, on everything from cures for childlessness to matters of war. Lavish gifts allowed the temple area to be built and decorated. Many states supplied their own treasury halls, where especially beautiful gifts were on display. Sideshows included the Omphalos, a stone reputedly marking the centre of the Earth, and the Pythic games, second only to the Olympic ones in importance. Strife inside Greece eventually led to a decline in the power of the Oracle. A serious blow was dealt when the Oracle of Zeus Ammon (see below) declared Alexander the Great to be the son of Zeus. The Pythic Oracle lived on, however, and was only closed by a Roman Imperial decree in 390 {\sc ad}, at a time when Christianity was ruthlessly destroying any religious opposition. Pythia then had been at the service of man and Gods for a millenium and a half. The r\^ole of the Pythic Oracle replies on the course of history is nowhere better described than in `The Histories' by Herodotus \cite{HerBC}, the classical and captivating description of the Ancient World at the time of the Great War between Greeks and Persians. Especially famous is the episode with King Croisus of Lydia. Contemplating a war against the upstart Persian Empire, he resolves to ask an oracle what the outcome of a potential battle would be. However, to have some guarantee for the veracity of any prophecy, he decides to send embassies to all the renowned oracles of the known World. The messengers are instructed to inquire the various divinities, on the hundredth day after their departure, what King Croisus is doing at that very moment. From the Pythia the messengers bring back the reply \begin{em}\begin{verse} I know the number of grains of sand as well as the expanse of the sea, \\ And I comprehend the dumb and hear him who does not speak, \\ There came to my mind the smell of the hard-shelled turtle, \\ Boiled in copper together with the lamb, \\ With copper below and copper above. \end{verse}\end{em} The veracity of the Pythia is thus established by the crafty ruler, who had waited until the appointed day, slaughtered a turtle and a lamb, and boiled them together in a copper cauldron with a copper lid. Also the Oracle of Zeus Ammon in the Libyan desert is able to give a correct reply (lost to posterity), while all others fail. King Croisus now sends a second embassy to Delphi, inquiring after the outcome of a battle against the Persians. The Pythia answers \begin{em}\begin{verse} If Croisus passes over the Halys he will dissolve a great Empire. \end{verse}\end{em} Taking this to mean he would win, the King collects his army and crosses the border river, only to suffer a crushing defeat and see his Kingdom conquered. When the victorious King Cyrus allows Croisus to send an embassy to upbraid the Oracle, the God Apollon answers through his Prophetess that he has correctly predicted the destruction of a great empire --- Croisus' own --- and that he cannot be held responsible if people choose to interpret the Oracle answers to their own liking. The history of the {\Py} program is neither as long nor as dignified as that of its eponym. However, some points of contact exist. You must be very careful when you formulate the questions: any ambiguities will corrupt the reply you get. And you must be even more careful not to misinterpret the answers; in particular not to pick the interpretation that suits you before considering the alternatives. Finally, even a perfect God has servants that are only human: a priest might mishear the screams of the Pythia and therefore produce an erroneous oracle reply; the current author might unwittingly let a bug free in the program {\Py}. \clearpage \section{Physics Overview} In this section we will try to give an overview of the main physics features of {\Je} and {\Py}, and also to introduce some terminology. The details will be discussed in subsequent sections. For the description of a typical high-energy event, an event generator should contain a simulation of several physics aspects. If we try to follow the evolution of an event in some semblance of a time order, one may arrange these aspects as follows: \begin{Enumerate} \item Initially two beam particles are coming in towards each other. Normally each particle is characterized by a set of parton distribution functions, which defines the partonic substructure in terms of flavour composition and energy sharing. \item One shower initiator parton from each beam starts off a sequence of branchings, such as $\q \to \q \g$, which build up an initial-state shower. \item One incoming parton from each of the two showers enters the hard process, where then a number of outgoing partons are produced, usually two. It is the nature of this process that determines the main characteristics of the event. \item Also the outgoing partons may branch, to build up final-state showers. \item When a shower initiator is taken out of a beam particle, a beam remnant is left behind. This remnant may have an internal structure, and a net colour charge that relates it to the rest of the final state. \item The QCD confinement mechanism ensures that the outgoing quarks and gluons are not observable, but instead fragment to colour neutral hadrons. \item Many of the produced hadrons are unstable and decay further. \end{Enumerate} Conventionally, only quarks and gluons are counted as partons, while leptons and photons are not. If pushed {\it ad absurdum} this may lead to some unwieldy terminology. We will therefore, where it does not matter, speak of an electron or a photon in the `partonic' substructure of an electron, lump branchings $\e \to \e \gamma$ together with other `parton shower' branchings such as $\q \to \q \g$, and so on. With this notation, the division into the above seven points applies equally well to an interaction between two leptons, between a lepton and a hadron, and between two hadrons. In the following subsections, we will survey the above seven aspects, not in the same order as given here, but rather in the order in which they appear in the program execution, i.e. starting with the hard process. \subsection{Hard Processes and Structure Functions} In {\Je}, only two hard processes are available. The first and main one is $\ee \to \gammaZ \to \q \qbar$. Here the `$*$' of $\gamma^*$ is used to denote that the photon must be off the mass shell. The distinction is of some importance, since a photon on the mass shell cannot decay. Of course also the $\Z^0$ can be off the mass shell, but here the distinction is less relevant (strictly speaking, a $\Z^0$ is always off the mass shell). In the following we may not always use `$*$' consistently, but the rule of thumb is to use a `$*$' only when a process is not kinematically possible for a particle of nominal mass. The quark $\q$ in the final state of $\ee \to \gammaZ \to \q \qbar$ may be $\u$, $\d$, $\s$, $\c$, $\b$ or $\t$; the flavour in each event is picked at random, according to the relative couplings, evaluated at the hadronic c.m. energy. Also the angular distribution of the final $\q \qbar$ pair is included. No parton-distribution functions are needed. The other {\Je} process is a routine to generate $\g \g \g$ and $\gamma \g \g$ final states, as expected in onium 1$^{--}$ decays such as $\Upsilon$. Given the current limits on the top mass, toponium will decay weakly much too fast for these processes to be of any interest, so therefore no new applications are expected. {\Py} contains a much richer selection, with close to a hundred different hard processes. These may be classified in many different ways. One is according to the number of final-state objects: we speak of `$2 \to 1$' processes, `$2 \to 2$' ones, `$2 \to 3$' ones, etc. This aspect is very relevant from a programming point of view: the more particles in the final state, the more complicated the phase space and therefore the whole generation procedure. In fact, {\Py} is optimized for $2 \to 1$ and $2 \to 2$ processes. There is currently no generic treatment of processes with three or more particles in the final state, but rather a few different machineries, each tailored to the pole structure of a specific class of graphs. This may be seen as a major limitation, and indeed is so at times. However, often one can come quite far with only one or two particles in the final state, since showers will add the required extra activity. The classification may also be misleading at times, since an $s$-channel resonance is considered as a single particle, even if it is assumed always to decay into two final-state particles. Thus the process $\ee \to \W^+ \W^- \to \q_1 \qbar'_1 \, \q_2 \qbar'_2$ is classified as $2 \to 2$, although the decay treatment of the $\W$ pair includes the full $2 \to 4$ matrix elements. Another classification is according to the physics scenario. This will be the main theme of section \ref{s:pytproc}. The following major groups may be distinguished: \begin{Itemize} \item Hard QCD processes, e.g. $\q \g \to \q \g$. \item Soft QCD processes, such as diffractive and elastic scattering, and minimum-bias events. \item Heavy-flavour production, e.g. $\g \g \to \t \tbar$. \item Prompt-photon production, e.g. $\q \g \to \q \gamma$. \item Photon-induced processes, e.g. $\gamma \g \to \q \qbar$. \item Deep inelastic scattering, e.g. $\q \ell \to \q \ell$. \item $\W / \Z$ production, such as the $\ee \to \gammaZ$ already found in {\Je}, or $\q \qbar \to \W^+ \W^-$. \item Standard model Higgs production, where the Higgs is reasonably light and narrow, and can therefore still be considered as a resonance. \item Gauge boson scattering processes, such as $\W \W \to \W \W$, when the Standard Model Higgs is so heavy and broad that resonant and non-resonant contributions have to be considered together. \item Non-standard Higgs particle production, within the framework of a two-Higgs-doublet scenario with three neutral and two charged Higgs states. \item Production of new gauge bosons, such as a $\Z'$. \item Production of fourth-generation fermions. \item Leptoquark production. \item Deviations from Standard Model processes, e.g. due to contact interactions or a strongly interacting gauge boson sector. These scenarios do not always appear as separate processes, but may just be options to some of the processes above. \end{Itemize} This is by no means a survey of all interesting physics. Most notable is the absence of supersymmetric particle production and decay, but many other examples could be found. Also, within the scenarios studied, not all contributing graphs have always been included, but only the more important and/or more interesting ones. In many cases, various approximations are involved in the matrix elements coded. The cross section for a given process $ij \to k$ is given by \begin{equation} \sigma_{ij \to k} = \int \d x_1 \int \d x_2 \, f^1_i(x_1) \, f^2_j(x_2) \, \hat{\sigma}_{ij \to k} ~. \end{equation} Here $\hat{\sigma}$ is the cross section for the hard partonic process, as codified in the matrix elements for each specific process. For processes with many particles in the final state it would be replaced by an integral over the allowed final-state phase space. The $f^a_i(x)$ are the parton-distribution functions, which describe the probability to find a parton $i$ inside beam particle $a$, with parton $i$ carrying a fraction $x$ of the total $a$ momentum. Actually, parton distributions also depend on some momentum scale $Q^2$ that characterizes the hard process. Structure functions are most familiar for hadrons, such as the proton. Hadrons are inherently composite objects, made up of quarks and gluons. Since we do not understand QCD, a derivation from first principles of hadron parton distributions does not yet exist, although some progress is being made in lattice QCD studies. It is therefore necessary to rely on parametrizations, where experimental data are used in conjunction with the evolution equations for the $Q^2$ dependence, to pin down the parton distributions. Several different groups have therefore produced their own fits, based on slightly different sets of data, and with some variation in the theoretical assumptions. Also for fundamental particles, such as the electron, is it convenient to introduce parton distributions. The function $f^{\e}_{\e}(x)$ thus parametrizes the probability that the electron that takes part in the hard process retains a fraction $x$ of the original energy, the rest being radiated (into photons) in the initial state. Of course, such radiation could equally well be made part of the hard interaction, but the parton-distribution approach usually is much more convenient. If need be, a description with fundamental electrons is recovered for the choice $f_{\e}^{\e}(x, Q^2)=\delta(x-1)$. Note that, contrary to the proton case, electron parton distributions are calculable from first principles, and reduce to the $\delta$ function above for $Q^2 \to 0$. The electron may also contain photons, and the photon may in its turn contain quarks and gluons. The internal structure of the photon is a bit of a problem, since the photon contains a point-like part, which is perturbatively calculable, and a vector-meson dominance part, which is not. Normally, the photon parton distributions are therefore parametrized, just as the hadron ones. Since the electron ultimately contains quarks and gluons, hard QCD processes like $\q \g \to \q \g$ therefore not only appear in $\pp$ collisions, but also in $\ep$ ones (`resolved photoproduction') and in $\ee$ ones (`doubly resolved 2$\gamma$ events'). The parton distribution function approach here makes it much easier to reuse one and the same hard process in different contexts. There is also another kind of possible generalization. The two processes $\q \qbar \to \gammaZ$, studied in hadron colliders, and $\ee \to \gammaZ$, studied in $\ee$ colliders, are really special cases of a common process, $\f \fbar \to \gammaZ$, where $\f$ denotes a fundamental fermion, i.e. a quark, lepton or neutrino. The whole structure is therefore only coded once, and then slightly different couplings and colour prefactors are used, depending on the initial state considered. Usually the interesting cross section is a sum over several different initial states, e.g. $\u \ubar \to \gammaZ$ and $\d \dbar \to \gammaZ$ in a hadron collider. This kind of summation is always implicitly done, even when not explicitly mentioned in the text. \subsection{Initial- and Final-State Radiation} In every process that contains coloured and/or charged objects in the initial or final state, gluon and/or photon radiation may give large corrections to the overall topology of events. Starting from a basic $2 \to 2$ process, this kind of corrections will generate $2 \to 3$, $2 \to 4$, and so on, final-state topologies. As the available energies are increased, hard emission of this kind is increasingly important, relative to fragmentation, in determining the event structure. Two traditional approaches exist to the modelling of perturbative corrections. One is the matrix-element method, in which Feynman diagrams are calculated, order by order. In principle, this is the correct approach, which takes into account exact kinematics, and the full interference and helicity structure. The only problem is that calculations become increasingly difficult in higher orders, in particular for the loop graphs. Only in exceptional cases have therefore more than one loop been calculated in full, and often we do not have any loop corrections at all at our disposal. On the other hand, we have indirect but strong evidence that, in fact, the emission of multiple soft gluons plays a significant r\^ole in building up the event structure, e.g. at LEP, and this sets a limit to the applicability of matrix elements. Since the phase space available for gluon emission increases with the available energy, the matrix-element approach becomes less relevant for the full structure of events at higher energies. However, the perturbative expansion by itself is better behaved at higher energies, owing to the running of $\alphas$. As a consequence, inclusive measurements, e.g. of the rate of well-separated jets, should yield more reliable results. The second possible approach is the parton-shower one. Here an arbitrary number of branchings of one parton into two (or more) may be put together, to yield a description of multijet events, with no explicit upper limit on the number of partons involved. This is possible since the full matrix-element expressions are not used, but only approximations derived by simplifying the kinematics, and the interference and helicity structure. Parton showers are therefore expected to give a good description of the substructure of jets, but in principle the shower approach has limited predictive power for the rate of well-separated jets (i.e. the 2/3/4/5-jet composition). In practice, shower programs may be patched up to describe the hard-gluon emission region reasonably well, in particular for the $\ee$ annihilation process. Nevertheless, the shower description is not optimal for absolute $\alphas$ determinations. Thus the two approaches are complementary in many respects, and both have found use. However, because of its simplicity and flexibility, the parton-shower option is generally the first choice, while the matrix elements one is mainly used for $\alphas$ determinations, angular distribution of jets, triple-gluon vertex studies, and other specialized studies. Obviously, the ultimate goal would be to have an approach where the best aspects of the two worlds are harmoniously married. \subsubsection{Matrix elements} Matrix elements are especially made use of in the {\Je} implementation of the process $\ee \to \gammaZ \to \q \qbar$. For initial-state QED radiation, a first order (unexponentiated) description has been adopted. This means that events are subdivided into two classes, those where a photon is radiated above some minimum energy, and those without such a photon. In the latter class, the soft and virtual corrections have been lumped together to give a total event rate that is correct up to one loop. This approach worked fine at PETRA/PEP energies, but does not do so well for the $\Z^0$ line shape, i.e. in regions where the cross section is rapidly varying and high precision is strived for. For final-state QCD radiation, several options are available. The default is the parton-shower one (see below), but the matrix-elements options are also frequently used. In the definition of 3- or 4-jet events, a cut is introduced whereby it is required that any two partons have an invariant mass bigger than some fraction of the c.m. energy. 3-jet events which do not fulfill this requirement are lumped with the 2-jet ones. The first-order matrix-element option, which only contains 3- and 2-jet events therefore involves no ambiguities. In second order, where also 4-jets have to be considered, a main issue is what to do with 4-jet events that fail the cuts. Depending on the choice of recombination scheme, whereby the two nearby partons are joined into one, different 3-jet events are produced. Therefore the second-order differential 3-jet rate has been the subject of some controversy, and {\Je} actually contains two different implementations. By contrast, {\Py} does not contain any full higher-order matrix elements, with loop contributions included. There are a few cases where higher-order matrix elements are included at the Born level. Consider e.g. the case of $\W$ production at a hadron collider, which is contained in the lowest-order process $\q \qbar' \to \W$. In an inclusive description, additional jets recoiling against the $\W$ may be generated by parton showers. {\Py} also contains the two first-order processes $\q \g \to \W \q'$ and $\q \qbar' \to \W \g$. The cross sections for these processes are divergent when the $\pT \to 0$. In this region a correct treatment would therefore have to take into account loop corrections, which are not available in {\Py}. Depending on the physics application, one could then use {\Py} in one of two ways. In the region of small $\pT$, the preferred option is lowest-order matrix elements combined with parton showers. For the production of a $\W$ at large $\pT$, on the other hand, the shower approach is too imprecise to give the right cross section; additionally the event selection machinery is very inefficient. Here it is advantageous to generate first-order events, and then add showers only to describe additional softer radiation. \subsubsection{Parton showers} The separation of radiation into initial- and final-state showers is arbitrary, but very convenient. There are also situations where it is appropriate: for instance, the process $\ee \to \Z^0 \to \q \qbar$ only contains final-state QCD radiation (QED radiation, however, is possible both in the initial and final state), while $\q \qbar \to \Z^0 \to \ee$ only contains initial-state QCD one. Similarly, the distinction of emission as coming either from the $\q$ or from the $\qbar$ is arbitrary. In general, the assignment of radiation to a given mother parton is a good approximation for an emission close to the direction of motion of that parton, but not for the wide-angle emission in between two jets, where interference terms are expected to be important. In both initial- and final-state showers, the structure is given in terms of branchings $a \to bc$, specifically $\e \to \e \gamma$, $\q \to \q \g$, $\q \to \q \gamma$, $\g \to \g \g$, and $\g \to \q \qbar$. Each of these processes is characterized by a splitting kernel $P_{a \to bc}(z)$. The branching rate is proportional to the integral $\int P_{a \to bc}(z) \, \d z$. The $z$ value picked for a branching describes the energy sharing, with daughter $b$ taking a fraction $z$ and daughter $c$ the remaining $1-z$ of the $a$ energy. Once formed, the daughters $b$ and $c$ may in turn branch, and so on. Each parton is characterized by some virtuality scale $Q^2$, which gives an approximate sense of time ordering to the cascade. In the initial-state shower, $Q^2$ values are gradually increasing as the hard scattering is approached, while $Q^2$ is decreasing in the final-state showers. Shower evolution is cut off at some lower scale $Q_0$, typically around 1 GeV for QCD branchings. The same cut-off scale is also used to regularize the soft gluon emission divergences in the splitting kernels. From above, a maximum scale $Q_{\mmax}$ is introduced, where the showers are matched to the hard interaction itself. The relation between $Q_{\mmax}$ and the kinematics of the hard scattering is uncertain, and the choice made can strongly affect the amount of well-separated jets. Despite a number of common traits, the initial- and final-state radiation machineries are in fact quite different, and are described separately below. For historical reasons, the final-state shower is found in {\Je} and the initial-state one in {\Py}. Final-state showers are time-like, i.e. partons have $m^2 = E^2 - \mbf{p}^2 \geq 0$. The evolution variable $Q^2$ of the cascade is therefore in {\Je} associated with the $m^2$ of the branching parton, but this choice is not unique. Starting from $Q^2_{\mmax}$, an original parton is evolved downwards in $Q^2$ until a branching occurs. The selected $Q^2$ value defines the mass of the branching parton, and the $z$ of the splitting kernel the parton energy division between its daughters. These daughters may now, in turn, evolve downwards, in this case with maximum virtuality already defined by kinematics, and so on down to the $Q_0$ cut-off. In QCD showers, corrections to the leading-log picture, so-called coherence effects, lead to an ordering of subsequent emissions in terms of decreasing angles. This does not follow automatically from the mass-ordering constraint, but is implemented as an additional requirement on allowed emissions. Photon emission is not affected by angular ordering. It is also possible to obtain non-trivial correlations between azimuthal angles in the various branchings, some of which are implemented as options. Finally, the theoretical analysis strongly suggests the scale choice $\alphas = \alphas(\pT^2) = \alphas(z(1-z)m^2)$, and this is the default in the program. The final-state radiation machinery is applied in the c.m. frame of the hard scattering. The total energy and momentum of the hard-scattering subsystem is preserved, as is the direction of the outgoing partons (in that frame). In contrast to final-state showers, initial-state ones are space-like. This means that, in the sequence of branchings $a \to bc$ that lead up from the shower initiator to the hard interaction, particles $a$ and $b$ have $m^2 = E^2 - \mbf{p}^2 <0$. The `side branch' particle $c$, which does not participate in the hard scattering, may be on the mass shell, or have a time-like virtuality. In the latter case a time-like shower will evolve off it, rather like the final-state radiation described above. To first approximation, the evolution of the space-like main branch is characterized by the evolution variable $Q^2 = -m^2$, which is required to be strictly increasing along the shower, i.e. $Q_b^2 > Q_a^2$. Corrections to this picture have been calculated, but are basically absent in {\Py}. Initial-state radiation is handled within the backwards evolution scheme. In this approach, the choice of the hard scattering is based on the use of evolved parton distributions, which means that the inclusive effects of initial-state radiation are already included. What remains is therefore to construct the exclusive showers. This is done starting from the two incoming partons at the hard interaction, tracing the showers `backwards in time', back to the two shower initiators. In other words, given a parton $b$, one tries to find the parton $a$ that branched into $b$. The evolution in the Monte Carlo is therefore in terms of a sequence of decreasing space-like virtualities $Q^2$ and increasing momentum fractions $x$. Branchings on the two sides are interleaved in a common sequence of decreasing $Q^2$ values. In the above formalism, there is no real distinction between gluon and photon emission. Some of the details actually do differ, as will be explained in the full description. The initial- and final-state radiation shifts around the kinematics of the original hard interaction. In deep inelastic scattering, this means that the $x$ and $Q^2$ values that can be derived from the momentum of the scattered lepton do not agree with the values originally picked. In high-$\pT$ processes, it means that one no longer has two jets with opposite and compensating $\pT$, but more complicated topologies. Effects of any original kinematics selection cuts are therefore smeared out, an unfortunate side-effect of the parton-shower approach. \subsection{Beam Remnants} In a hadron--hadron collision, the initial-state radiation algorithm reconstructs one shower initiator in each beam. This initiator only takes some fraction of the total beam energy, leaving behind a beam remnant which takes the rest. For a proton beam, a $\u$ quark initiator would leave behind a $\u \d$ diquark beam remnant, with an antitriplet colour charge. The remnant is therefore colour-connected to the hard interaction, and forms part of the same fragmenting system. It is further customary to assign a primordial transverse momentum to the shower initiator, to take into account the motion of quarks inside the original hadron, basically as required by the uncertainty principle. This primordial $k_{\perp}$ is selected according to some suitable distribution, and the recoil is assumed to be taken up by the beam remnant. Often the remnant is more complicated, e.g. a $\g$ initiator would leave behind a $\u \u \d$ proton remnant system in a colour octet state, which can conveniently be subdivided into a colour triplet quark and a colour antitriplet diquark, each of which are colour-connected to the hard interaction. The energy sharing between these two remnant objects, and their relative transverse momentum, introduces additional degrees of freedom, which are not understood from first principles. Na\"{\i}vely, one would expect an $\ep$ event to have only one beam remnant, and an $\ee$ event none. This is not always correct, e.g. a $\gamma \gamma \to \q \qbar$ interaction in an $\ee$ event would leave behind the $\e^+$ and $\e^-$ as beam remnants, and a $\q \qbar \to \g \g$ interaction in resolved photoproduction in an $\ee$ event would leave behind one $\e^{\pm}$ and one $\q / \qbar$ in each remnant. Corresponding complications occur for photoproduction in $\ep$ events. There is another source of beam remnants. If parton distributions are used to resolve an electron inside an electron, some of the original energy is not used in the hard interaction, but is rather associated with initial-state photon radiation. The initial-state shower is in principle intended to trace this evolution and reconstruct the original electron before any radiation at all took place. However, because of cut-off procedures, some small amount may be left unaccounted. Alternatively the user may have chosen to switch off initial-state radiation altogether, but still preserved the resolved electron parton distributions. In either case the remaining energy is given to a single photon of vanishing transverse momentum, which is then considered in the same spirit as `true' beam remnants. So far we have assumed that each event only contains one hard interaction, i.e. that each incoming particle has only one parton which takes part in hard processes, and that all other constituents sail through unaffected. This is appropriate in $\ee$ or $\ep$ events, but not necessarily so in hadron--hadron collisions. Here each of the beam particles contains a multitude of partons, and so the probability for several interactions in one and the same event need not be negligible. In principle these additional interactions could arise because one single parton from one beam scatters against several different partons from the other beam, or because several partons from each beam take place in separate $2 \to 2$ scatterings. Both are expected, but combinatorics should favour the latter, which is the mechanism considered in {\Py}. The dominant $2 \to 2$ QCD cross sections are divergent for $\pT \to 0$, and drop rapidly for larger $\pT$. Probably the lowest-order perturbative cross sections will be regularized at small $\pT$ by colour coherence effects: an exchanged gluon of small $\pT$ has a large transverse wave function and can therefore not resolve the individual colour charges of the two incoming hadrons; it will only couple to an average colour charge that vanishes in the limit $\pT \to 0$. In the program, some effective $\pTmin$ scale is therefore introduced, below which the perturbative cross section is either assumed completely vanishing or at least strongly damped. Phenomenologically, $\pTmin$ comes out to be a number of the order of 1.5--2.0 GeV. In a typical `minimum-bias' event one therefore expects to find one or a few scatterings at scales around or a bit above $\pTmin$, while a high-$\pT$ event also may have additional scatterings at the $\pTmin$ scale. The probability to have several high-$\pT$ scatterings in the same event is small, since the cross section drops so rapidly with $\pT$. The understanding of multiple interaction is still very primitive, and even the experimental evidence that it exists at all is rather weak. {\Py} therefore contains several different options, with a fairly simple one as default. The options differ in particular on the issue of the `pedestal' effect: is there an increased probability or not for additional interactions in an event which is known to contain a hard scattering, compared with one that contains no hard interactions? \subsection{Fragmentation} QCD perturbation theory, formulated in terms of quarks and gluons, is valid at short distances. At long distances, QCD becomes strongly interacting and perturbation theory breaks down. In this confinement regime, the coloured partons are transformed into colourless hadrons, a process called either hadronization or fragmentation. In this paper we reserve the former term for the combination of fragmentation and the subsequent decay of unstable particles. The fragmentation process has yet to be understood from first principles, starting from the QCD Lagrangian. This has left the way clear for the development of a number of different phenomenological models. Three main schools are usually distinguished, string fragmentation (SF), independent fragmentation (IF) and cluster fragmentation (CF), but many variants and hybrids exist. Being models, none of them can lay claims to being `correct', although some may be better founded than others. The best that can be aimed for is internal consistency, a good representation of existing data, and a predictive power for properties not yet studied or results at higher energies. {\Je} is intimately connected with string fragmentation, in the form of the time-honoured `Lund model'. This is the default for all {\JePy} applications, but independent fragmentation options also exist, for applications where one wishes to study the importance of string effects. All current models are of a probabilistic and iterative nature. This means that the fragmentation process as a whole is described in terms of one or a few simple underlying branchings, of the type jet $\to$ hadron + remainder-jet, string $\to$ hadron + remainder-string, and so on. At each branching, probabilistic rules are given for the production of new flavours, and for the sharing of energy and momentum between the products. To understand fragmentation models, it is useful to start with the simplest possible system, a colour-singlet $\q \qbar$ 2-jet event, as produced in $\ee$ annihilation. Here lattice QCD studies lend support to a linear confinement picture (in the absence of dynamical quarks), i.e. the energy stored in the colour dipole field between a charge and an anticharge increases linearly with the separation between the charges, if the short-distance Coulomb term is neglected. This is quite different from the behaviour in QED, and is related to the presence of a triple-gluon vertex in QCD. The details are not yet well understood, however. The assumption of linear confinement provides the starting point for the string model. As the $\q$ and $\qbar$ partons move apart from their common production vertex, the physical picture is that of a colour flux tube (or maybe colour vortex line) being stretched between the $\q$ and the $\qbar$. The transverse dimensions of the tube are of typical hadronic sizes, roughly 1 fm. If the tube is assumed to be uniform along its length, this automatically leads to a confinement picture with a linearly rising potential. In order to obtain a Lorentz covariant and causal description of the energy flow due to this linear confinement, the most straightforward way is to use the dynamics of the massless relativistic string with no transverse degrees of freedom. The mathematical, one-dimensional string can be thought of as parametrizing the position of the axis of a cylindrically symmetric flux tube. From hadron spectroscopy, the string constant, i.e. the amount of energy per unit length, is deduced to be $ \kappa \approx 1$ GeV/fm. The expression `massless' relativistic string is somewhat of a misnomer: $\kappa$ effectively corresponds to a `mass density' along the string. Let us now turn to the fragmentation process. As the $\q$ and $\qbar$ move apart, the potential energy stored in the string increases, and the string may break by the production of a new $\q' \qbar'$ pair, so that the system splits into two colour-singlet systems $\q \qbar'$ and $\q' \qbar$. If the invariant mass of either of these string pieces is large enough, further breaks may occur. In the Lund string model, the string break-up process is assumed to proceed until only on-mass-shell hadrons remain, each hadron corresponding to a small piece of string with a quark in one end and an antiquark in the other. In order to generate the quark--antiquark pairs $\q' \qbar'$ which lead to string break-ups, the Lund model invokes the idea of quantum mechanical tunnelling. This leads to a flavour-independent Gaussian spectrum for the $\pT$ of $\q' \qbar'$ pairs. Since the string is assumed to have no transverse excitations, this $\pT$ is locally compensated between the quark and the antiquark of the pair. The total $\pT$ of a hadron is made up out of the $\pT$ contributions from the quark and antiquark that together form the hadron. Some contribution of very soft perturbative gluon emission may also effectively be included in this description. The tunnelling picture also implies a suppression of heavy-quark production, $\u : \d : \s : \c \approx 1 : 1 : 0.3 : 10^{-11}$. Charm and heavier quarks hence are not expected to be produced in the soft fragmentation, but only in perturbative parton-shower branchings $\g \to \q \qbar$. When the quark and antiquark from two adjacent string breakings are combined to form a meson, it is necessary to invoke an algorithm to choose between the different allowed possibilities, notably between pseudoscalar and vector mesons. Here the string model is not particularly predictive. Qualitatively one expects a $1 : 3$ ratio, from counting the number of spin states, multiplied by some wave-function normalization factor, which should disfavour heavier states. A tunnelling mechanism can also be used to explain the production of baryons. This is still a poorly understood area. In the simplest possible approach, a diquark in a colour antitriplet state is just treated like an ordinary antiquark, such that a string can break either by quark--antiquark or antidiquark--diquark pair production. A more complex scenario is the `popcorn' one, where diquarks as such do not exist, but rather quark--antiquark pairs are produced one after the other. This latter picture gives a less strong correlation in flavour and momentum space between the baryon and the antibaryon of a pair. In general, the different string breaks are causally disconnected. This means that it is possible to describe the breaks in any convenient order, e.g. from the quark end inwards. One therefore is led to write down an iterative scheme for the fragmentation, as follows. Assume an initial quark $\q$ moving out along the $+z$ axis, with the antiquark going out in the opposite direction. By the production of a $\q_1 \qbar_1$ pair, a meson $\q \qbar_1$ is produced, leaving behind an unpaired quark $\q_1$. A second pair $\q_2 \qbar_2$ may now be produced, to give a new meson $\q_1 \qbar_2$, etc. At each step the produced hadron takes some fraction of the available energy and momentum. This process may be iterated until all energy is used up, with some modifications close to the $\qbar$ end of the string in order to make total energy and momentum come out right. The choice of starting the fragmentation from the quark end is arbitrary, however. A fragmentation process described in terms of starting at the $\qbar$ end of the system and fragmenting towards the $\q$ end should be equivalent. This `left--right' symmetry constrains the allowed shape of the fragmentation function $f(z)$, where $z$ is the fraction of the remaining light-cone momentum $E \pm p_z$ (+ for the $\q$ jet, $-$ for the $\qbar$ one) taken by each new particle. The resulting `Lund symmetric fragmentation function' has two free parameters, which are determined from data. If several partons are moving apart from a common origin, the details of the string drawing become more complicated. For a $\q \qbar \g$ event, a string is stretched from the $\q$ end via the $\g$ to the $\qbar$ end, i.e. the gluon is a kink on the string, carrying energy and momentum. As a consequence, the gluon has two string pieces attached, and the ratio of gluon to quark string force is 2, a number which can be compared with the ratio of colour charge Casimir operators, $N_C/C_F = 2/(1-1/N_C^2) = 9/4$. In this, as in other respects, the string model can be viewed as a variant of QCD where the number of colours $N_C$ is not 3 but infinite. Note that the factor 2 above does not depend on the kinematical configuration: a smaller opening angle between two partons corresponds to a smaller string length drawn out per unit time, but also to an increased transverse velocity of the string piece, which gives an exactly compensating boost factor in the energy density per unit string length. The $\q \qbar \g$ string will fragment along its length. To first approximation this means that there is one fragmenting string piece between $\q$ and $\g$ and a second one between $\g$ and $\qbar$. One hadron is straddling both string pieces, i.e. sitting around the gluon corner. The rest of the particles are produced as in two simple $\q \qbar$ strings, but strings boosted with respect to the overall c.m. frame. When considered in detail, the string motion and fragmentation is more complicated, with the appearance of additional string regions during the time evolution of the system. These corrections are especially important for soft and collinear gluons, since they provide a smooth transition between events where such radiation took place and events where it did not. Therefore the string fragmentation scheme is `infrared safe' with respect to soft or collinear gluon emission. For events that involve many partons, there may be several possible topologies for their ordering along the string. An example would be a $\q \qbar \g_1 \g_2$ (the gluon indices are here used to label two different gluon-momentum vectors), where the string can connect the partons in either of the sequences $\q - \g_1 - \g_2 - \qbar$ and $\q - \g_2 - \g_1 - \qbar$. The matrix elements that are calculable in perturbation theory contain interference terms between these two possibilities, which means that the colour flow is not always well-defined. Fortunately, the interference terms are down in magnitude by a factor $1/N_C^2$, where $N_C = 3$ is the number of colours, so approximate recipes can be found. In the leading log shower description, on the other hand, the rules for the colour flow are well-defined. A final comment: in the argumentation for the importance of colour flows there is a tacit assumption that soft-gluon exchanges between partons will not normally mess up the original colour assignment; this is likely the case but has not been proven. \subsection{Decays} A large fraction of the particles produced by fragmentation are unstable and subsequently decay into the observable stable (or almost stable) ones. It is therefore important to include all particles with their proper mass distributions and decay properties. Although involving little deep physics, this is less trivial than it may sound: while a lot of experimental information is available, there is also very much that is missing. For charm mesons, it is necessary to put together measured exclusive branching ratios with some inclusive multiplicity distributions to obtain a consistent and reasonably complete set of decay channels, a rather delicate task. For bottom, so far only a rather simple phase-space type of generator has been used for hadronic decays. Normally it is assumed that decay products are distributed according to phase space, i.e. that there is no dynamics involved in their relative distribution. However, in many cases additional assumptions are necessary, e.g. for semileptonic decays of charm and bottom hadrons one needs to include the proper weak matrix elements. Particles may also be produced polarized and impart a non-isotropic distribution to their decay products. Many of these effects are not at all treated in the program. In fact, spin information is not at all carried along, but has to be reconstructed explicitly when needed. The normal decay treatment is handled by {\Je}, making use of a set of tables where branching ratios and decay modes are stored. In {\Py} a separate decay treatment exists, used exclusively for a specific list of particles: $\Z^0$, $\W^{\pm}$, $\H^0$, $\Z'^0$, $\W'^{\pm}$, $\H'^0$, $\A^0$, $\H^{\pm}$, $\eta_{\mrm{tech}}^0$, $\R^0$, $\q^*$, $\ell^*$, and the leptoquark $\L_{\Q}$. Together we call these resonances, and contrast the `particle decay' treatment of {\Je} with the `resonance decay' one of {\Py}. Of course, this is just a matter of terminology: a particle like the $\rho$ could also be called a resonance. What characterizes a ({\Py}) resonance is that partial widths and branching ratios are calculated dynamically, as a function of the actual mass of a particle. Therefore not only do branching ratios change between an $\H^0$ of nominal mass 100 GeV and one of 200 GeV, but also for a Higgs of nominal mass 200 GeV, the branching ratios would change between an actual mass of 190 GeV and 210 GeV, say. This is particularly relevant for reasonably broad resonances, and in threshold regions. For an approach like this to work, it is clearly necessary to have perturbative expressions available for all partial widths, which is one reason why a corresponding treatment would not be the same for an ordinary hadronic resonance, like the $\rho$. The decay products of {\Py} resonances are typically quarks, leptons, or other resonances, e.g. $\W \to \q \qbar'$ or $\H^0 \to \W^+ \W^-$. In decays to quarks, parton showers are automatically added to give a more realistic multijet structure, and one may also allow photon emission off leptons. If the decay products in turn are resonances, further decays are necessary. Often spin information is available in resonance decay matrix elements, contrary to the normal state of affairs in ordinary particle decays. This means that the angular orientations in the two decays of a $\W^+ \W^-$ pair are properly correlated. Occasionally, the information is not available, and then resonances decay isotropically. The top quark is a special problem. The original machinery is based on the assumption that the $\t$ is long-lived, so that top hadrons have time to form in the fragmentation process, and afterwards these mesons decay weakly. With current `best bet' mass values, this is not correct, but one should rather consider top decay before fragmentation. Top should then be handled like one of the above resonances. Therefore the program now contains an alternative along these lines, which is the preferred option. \clearpage \section{Program Overview} This section contains a diverse collection of information. The first part is an overview of previous {\Je} and {\Py} versions. The second gives instructions for installation of the programs and describes their philosophy: how they are constructed and how they are supposed to be used. It also contains some information on how to read this manual. The third and final part contains several examples of pieces of code or short programs, to illustrate the general style of program usage. The last part is mainly intended as an introduction for completely new users, and can be skipped by more experienced ones. Since the {\Je} and {\Py} programs today are so closely connected, and are gradually coalescing, they are presented together in this report. However, they still appear as separate entities, with slightly different style and emphasis. {\Je} is the older of the two, and is at the origin of the whole `Lund' family of event generators. It can be subdivided in two parts. The larger is a generic package for jet fragmentation, particle decays, final-state parton showers, event-analysis routines, and other utitilies. This package can be used in the context of any hard process, provided one is willing to buy the underlying assumption of jet universality, i.e. that the fragmentation process is fundamentally the same whether one is considering an $\ee$ or a $\pp$ event, and that the only differences are to be found in the parton-level processes involved. This package is not only used by all other `Lund' programs, but also by numerous other programs written to study specific processes. The smaller part of {\Je} is a generator for $\ee$ annihilation events, according to either a parton-shower or a matrix-element approach. The {\Je} program is completely selfcontained. {\Py} is a program made to generate hard or soft processes in collisions between leptons, hadrons and photons, especially at $\ee$, $\ep$ and $\pp$ colliders. Where {\Je} is a loose collection of routines that you can combine as desired, {\Py} is a more structured program, where you initially set up what processes you want to study, and thereafter all events will be generated according to this specification. Included is an extensive library of hard subprocess differential cross sections, a library of parton distributions, a process generation machinery, treatment of initial-state showers and beam remnants, and a few odds and ends. {\Je} is used for final-state showers, fragmentation and decay, but no other external libraries are needed. An interface to external parton-distribution function libraries is provided, however. Many programs written by other persons make use of {\Je}, and a few also of {\Py}. It is not my intention to give a complete list here. A majority of these programs are specific to given collaborations, and therefore not publicly distributed. Below we give a list of a few public programs from the `Lund group', which may have a somewhat wider application. None of them are supported by the current author, so any requests should be directed to the persons mentioned. \begin{Itemize} \item {\sc Ariadne} is a generator for dipole emission, written mainly by L. L\"onnblad \cite{Pet88}. The dipole provides an alternative formulation of initial- and final-state showers. {\Je} or {\Py} can be used to generate the hard process and {\Je} to do the fragmentation. \item {\sc Aroma} is a generator for heavy-flavour processes in leptoproduction, written by G.~Ingelman and G. Schuler \cite{Ing88}. It uses {\Je} for fragmentation. \item {\sc Fritiof} is a generator for hadron--hadron, hadron--nucleus and nucleus--nucleus collisions \cite{Nil87}, which makes use of {\Py} to generate hard QCD scatterings and of {\Je} for fragmentation. Currently H. Pi is responsible for program development. \item {\sc Lepto} is a leptoproduction event generator, written mainly by G. Ingelman \cite{Ing80}. It can generate parton configurations in deep inelastic scattering according to a number of possibilities. It makes use of {\Je} for fragmentation and additionally has a parton-shower option based on {\Py}. \item {\sc Lucifer} is a photoproduction generator written by G. Ingelman and A. Weigend \cite{Ing87a}. It is a modification of an earlier version of {\Py} and makes use of {\Je}. \item {\sc Pompyt} is a generator for pomeron interactions written by P. Bruni and G. Ingelman \cite{Bru93}. This program defines parton distributions, flux factors and other aspects specific to the pomeron, which is combined with the standard {\Py} machinery for process generation. \item {\sc Twister} is a generator for higher-twist processes, written by G. Ingelman \cite{Ing87}. It is a modification of an earlier version of {\Py} and makes use of {\Je}. \end{Itemize} One should also note that a version of {\Py} has been modified to include the effects of longitudinally polarized incoming protons. This is the work of St. G\"ullenstern et al. \cite{Gul93}. \subsection{Update History} Both {\Je} and {\Py} are by now fairly old and well-established programs, but they are still steadily being improved on. While evolution was especially rapid for {\Je} in the early days, that program has by now reached a certain level of maturity, and the pace of change has dropped significantly. {\Py}, on the other hand, has been continually extended in recent years, and may still see further growth, although most of the basic structure should be in place by now. In earlier days, before the advent of electronic mail, programs were only infrequently distributed, and version numbers corresponded to distinct new upgrades. Today, the evolutionary process is more continuous and so is the distribution of new versions. In particular, the introduction of a new process or feature is often done on short notice, if no problems of backwards compatibility are involved. With this distribution, the subversion numbers have therefore been expanded to three digits, where the last two give sub-subversions. For every change made in the public file, the sub-subversion number is updated, together with the `last date of change'. In most referencing the shorter `{\Je} version 7.4' could still be preferable to e.g. `{\Je} version 7.412'. For the record, in Tables \ref{Jetsetver} and \ref{Pythiaver} we list the official main versions of {\Je} and {\Py}, respectively, with some brief comments. \begin{table}[tp] \captive{The main versions of {\Je}, with their date of appearance, published manuals, and main changes from previous versions. \label{Jetsetver}} \\ \vspace{1ex} \begin{center} \begin{tabular}{|c|c|c|l|@{\protect\rule{0mm}{\tablinsep}}} \hline No. & Date & Publ. & Main new or improved features \\[1mm] \hline 1 & Nov 78 & \cite{Sjo78} & single-quark jets \\ 2 & May 79 & \cite{Sjo79} & heavy-flavour jets \\ 3.1 & Aug 79 & --- & 2-jets in $\ee$, preliminary 3-jets \\ 3.2 & Apr 80 & \cite{Sjo80} & 3-jets in $\ee$ with full matrix elements, \\ & & & toponium $\to \g \g \g$ decays \\ 3.3 & Aug 80 & --- & softer fragmentation spectrum \\ 4.1 & Apr 81 & --- & baryon production and diquark fragmentation, \\ & & & fourth-generation quarks, larger jet systems \\ 4.2 & Nov 81 & --- & low-$\pT$ physics \\ 4.3 & Mar 82 & \cite{Sjo82} & 4-jets and QFD structure in $\ee$, \\ & Jul 82 & \cite{Sjo83} & event-analysis routines \\ 5.1 & Apr 83 & --- & improved string fragmentation scheme, symmetric \\ & & & fragmentation, full 2$^{\mrm{nd}}$ order QCD for $\ee$ \\ 5.2 & Nov 83 & --- & momentum-conservation schemes for IF, \\ & & & initial-state photon radiation in $\ee$ \\ 5.3 & May 84 & --- & `popcorn' model for baryon production \\ 6.1 & Jan 85 & --- & common blocks restructured, parton showers \\ 6.2 & Oct 85 & \cite{Sjo86} & error detection \\ 6.3 & Oct 86 & \cite{Sjo87} & new parton-shower scheme \\ 7.1 & Feb 89 & --- & new particle codes and common block structure, \\ & & & more mesons, improved decays, vertex information, \\ & & & Abelian gluon model, Bose--Einstein effects \\ 7.2 & Nov 89 & --- & interface to new standard common block, \\ & & & photon emission in showers \\ 7.3 & May 90 & \cite{Sjo92d} & expanded support for non-standard particles \\ 7.4 & Dec 93 & \cite{Sjo93} & updated particle data and defaults \\[1mm] \hline \end{tabular} \end{center} \end{table} \begin{table}[tp] \captive{The main versions of {\Py}, with their date of appearance, published manuals, and main changes from previous versions. \label{Pythiaver}} \\ \vspace{1ex} \begin{center} \begin{tabular}{|c|c|c|l|@{\protect\rule{0mm}{\tablinsep}}} \hline No. & Date & Publ. & Main new or improved features \\[1mm] \hline 1 & Dec 82 & \cite{Ben84} & synthesis of predecessors {\sc Compton}, {\sc Highpt} and \\ & & & {\sc Kassandra} \\ 2 & --- & & \\ 3.1 & --- & & \\ 3.2 & --- & & \\ 3.3 & Feb 84 & \cite{Ben84a} & scale-breaking parton distributions \\ 3.4 & Sep 84 & \cite{Ben85} & more efficient kinematics selection \\ 4.1 & Dec 84 & & initial- and final-state parton showers, $\W$ and $\Z$ \\ 4.2 & Jun 85 & & multiple interactions \\ 4.3 & Aug 85 & & $\W \W$, $\W \Z$, $\Z \Z$ and $\R$ processes \\ 4.4 & Nov 85 & & $\gamma \W$, $\gamma \Z$, $\gamma \gamma$ processes \\ 4.5 & Jan 86 & & $\H^0$ production, diffractive and elastic events \\ 4.6 & May 86 & & angular correlation in resonance pair decays \\ 4.7 & May 86 & & $\Z'^0$ and $\H^+$ processes \\ 4.8 & Jan 87 & \cite{Ben87} & variable impact parameter in multiple interactions \\ 4.9 & May 87 & & $\g \H^+$ process \\ 5.1 & May 87 & & massive matrix elements for heavy quarks \\ 5.2 & Jun 87 & & intermediate boson scattering \\ 5.3 & Oct 89 & & new particle and subprocess codes, new common block \\ & & & structure, new kinematics selection, some \\ & & & lepton--lepton and lepton--hadron interactions, \\ & & & new subprocesses \\ 5.4 & Jun 90 & & $s$-dependent widths, resonances not on the mass shell, \\ & & & new processes, new parton distributions \\ 5.5 & Jan 91 & & improved $\ee$ and $\ep$, several new processes \\ 5.6 & Sep 91 & \cite{Sjo92d} & reorganized parton distributions, new processes, \\ & & & user-defined external processes \\ 5.7 & Dec 93 & \cite{Sjo93} & new total cross sections, photoproduction, top decay \\[1mm] \hline \end{tabular} \end{center} \end{table} All versions preceding {\Je} 7.3 and {\Py} 5.6 should now be considered obsolete, and are no longer supported. For stable applications, the earlier combination {\Je} 6.3 and {\Py} 4.8 could still be used, however. {\Je} version 7 and {\Py} version 5 have been evolved in parallel, so some of the processes added in later versions of {\Py} make use of particle data only found in {\Je} from that time onwards. Although it would be possible to combine {\Py} 5.7 with {\Je} 7.3, e.g., it is not recommended. From the current versions onwards, checks have therefore been introduced to detect the use of (potentially) incompatible subversions, with warnings issued at initialization if that should be the case. Previous versions of the manuals have contained detailed lists of modifications from one version to the next, see e.g. \cite{Sjo92d}. Below we only reproduce the updates that appear with the most recent versions of the programs. Some of them were introduced in later editions of {\Py} 5.6 with {\Je} 7.3, while others are completely new. If nothing is explicitly said, these changes do not affect backwards compatibility, but only add new features. \subsubsection{Updates in JETSET 7.4} Changes from version 7.3 to 7.4 are not so large, although the impact of the updated particle data and parameter default values may need to be studied. \begin{Itemize} \item Particle data have been updated in accordance with the 1992 Review of Particle Properites \cite{PDG92}. (As usual, with a free interpretation of inconsistencies, unclarities and other gaps in the knowledge.) Changes are especially drastic for charm and bottom. In the bottom sector the decay properties are now given individually for $\B^0$, $\B^+$, $\B_s^0$, $\B_c^+$ and $\Lambda_b^0$, i.e. the generic data for `pseudoparticle' 85 are only used for other weakly decaying $\B$ baryons. \item Also a few other Standard Model parameters have been updated, such as the $\Z^0$ and $\W^{\pm}$ masses and widths, $\ssintw$ and the CKM matrix elements. \item Fragmentation and parton shower parameters have been modified to reflect current LEP knowledge \cite{LEP90}, i.e. a minor retuning starting from an average of the `best' parameter values obtained by the four LEP collaborations. Bose-Einstein effects are still left out. Flavour composition is unchanged, except for a suppression of $\eta'$ production. Affected by the change are {\tt MSTJ(11)}, {\tt PARJ(21)}, {\tt PARJ(23)}, {\tt PARJ(24)}, {\tt PARJ(41)}, {\tt PARJ(42)}, {\tt PARJ(54)}--{\tt PARJ(58)} and {\tt PARJ(81)}. \item Several other default values have been changed for switches and parameters in the $\ee$, parton shower and fragmentation parts of the programs. These changes are intended to reflect our current best knowledge. See {\tt MSTJ(26)}, {\tt MSTJ(27)}, {\tt MSTJ(41)}, {\tt MSTJ(46)}, {\tt MSTJ(50)}, {\tt MSTJ(110)}, {\tt PARJ(26)}, and {\tt PARJ(121)}--{\tt PARJ(125)}. \item A common title page for {\Je} and {\Py} has been introduced with the {\tt LULOGO} routine. Sub-subversion numbers are also given. \item Several options have been added for the {\tt LUSHOW} shower routine. See {\tt MSTJ(41)}, {\tt MSTJ(47)} and {\tt MSTJ(50)}. \item A $\b$ quark produced in the decay of a top hadron is allowed to radiate according to the standard parton shower scheme. \item The scalar gluon option contains the full electroweak angular distribution of 3-jet events. \item The {\tt LUCOMP} routine has been modified. Among other things, the $\B_c^+$ now appears as a separate compressed code, further codes for diffractive states have been added to the current list, and the pomeron (reggeon, $\eta_{\mrm{techni}}$) has been added as particle 29 (28, 38). \item A minimum threshold for calorimeter cell energy has been introduced for the {\tt LUCELL} routine. \item All obsolescent features of the Fortran 90 standard have been removed, i.e. the program should work well either with a Fortran 77 compiler or with a Fortran 90 one. \item A few minor errors have been corrected. \end{Itemize} %The following changes have been made since the beginning of 1994, %i.e. since the original distribution 7.400: %\begin{Itemize} % item %\end{Itemize} \subsubsection{Updates in PYTHIA 5.7} The updates from version 5.6 to 5.7 are all minor, and just about any program that ran with version 5.6 will also work with {\Py} 5.7. However, as for {\Je}, it should be noted that some important default values have been changed. \begin{Itemize} \item New parametrizations of the total cross sections of hadronic reactions, based in Donnachie--Landshoff \cite{Don92}, which replace the old ones. \item New parametrizations of elastic and single and double diffractive cross sections of hadronic reactions, based on Schuler--Sj\"ostrand \cite{Sch93b,Sch93a}, which replace the old ones. Also the slope parameters, the diffractive mass distributions and other aspects of the event generation have been changed accordingly. \item A possibility to give own total, elastic and diffractive cross sections. \item The single diffractive cross section has been split into its two constituents, $AB \to XB$ and $AB \to AX$. As a consequence, the diffractive subprocess codes 92--94 have received changed meaning. \item A new common block {\tt PYINT7} has been added for the expanded total cross section information, and this information has been partly removed from other common blocks. \item A much extended description of photoproduction physics, with the possibility to simulate separately VMD, anomalous and direct processes \cite{Sch93,Sch93a}. \item The selection of proton parton distributions that come with the program has been updated with the CTEQ2 ones, while some others have been removed. New default is the leading-order fit CTEQ2L. \item Since the {\sc Pdflib} library now has been expanded to contain also parton-distribution functions for the photon, the interfaces to the {\sc Pakpdf} and {\sc Phopdf} libraries have been removed. In addition, the interface to {\sc Pdflib} has been modified, and is now for appropriate for {\sc Pdflib} version 4. \item An extension of hadron parton distributions into the low-$x$ and low-$Q^2$ region \cite{Sch93a}. \item The top quark can be made to decay before it has time to fragment. In view of the current best estimate for the top mass, this is the expected behaviour, and is therefore now default. Further, a parton shower is allowed to evolve in the top decay. Also fourth generation quarks are allowed to decay before they fragment, and so on. \item It is possible to call {\tt PYEVNT} with energies that vary from one event to the next, without the need to reinitialize. \item Improved scheme for post-factor conservation of $x$ and $Q^2$ in deep inelastic scattering. \item Processes 15, 19, 30 and 35 have been expanded to cover $\gamma^*$ production in addition to the $\Z^0$ one, with full interference. \item New process 80, $\q\gamma \to \q'\pi^{\pm}$. \item New process 110, $\f\fbar \to \gamma\H^0$. \item New process 149, $\g\g \to \eta_{\mrm{techni}}$. \item New option for initial state radiation to restrict angular range of emission in accordance with coherence considerations. \item Some options have been added or removed, and default values have been changed. This includes {\tt KFIN} (top parton distributions off by default), {\tt MSTP(7)}, {\tt MSTP(11)}, {\tt MSTP(14)}, {\tt MSTP(23)}, {\tt MSTP(30)} (removed), {\tt MSTP(31)}, {\tt MSTP(34)}, {\tt MSTP(45)}, {\tt MSTP(48)}, {\tt MSTP(49)}, {\tt MSTP(62)}, {\tt MSTP(67)}, {\tt MSTP(101)}, {\tt PARP(13)}, {\tt PARP(81)}, {\tt PARP(82)}, {\tt PARP(47)} and {\tt PARP(101)}. \item All obsolescent features of the Fortran 90 standard have been removed, i.e. the program should work well either with a Fortran 77 compiler or with a Fortran 90 one. \item A few minor errors have been corrected. \end{Itemize} %The following changes have been made since the beginning of 1994, %i.e. since the original distribution 5.700: %\begin{Itemize} %\item %\end{Itemize} \subsection{Program Installation} \label{ss:install} Several `authorized' sources of the programs exist. The `master copy' of the programs is the one found on the {\tt TORSJO 192} disk on {\tt CERNVM}. There you have: \begin{tabbing} ~~~ \= {\tt JETSET74 FORTRAN}~~~ \= the {\Je} code, \\ \> {\tt PYTHIA57 FORTRAN} \> the {\Py} code, and \\ \> {\tt PYTHIA57 TEX} \> this common {\PyJe} manual. \end{tabbing} In addition to these, one may also find compiled {\tt TEXT} files, which are ready to be linked if you run jobs on the {\tt CERNVM} machine. The versions found here are updated without advance warning, as bugs are corrected or new features added. As new versions appear in the future ({\Je} 7.4 and {\Py} 5.7, etc.), the old program and manual copies on the disk are going to be removed. The lack of stable versions may make it less convenient to rely on the above files. New versions are introduced in the general distribution of the CERN program library, maybe once a year. These versions are better checked before release, and should be useful for most applications. However, clearly, they may be less up-to-date. Read the CERN Computer Newsletter for announcements. Copies of the programs are also available via anonymous ftp, e.g. from the asis01 server at CERN. The programs are written entirely in standard Fortran 77, and should run on any machine with such a compiler. To a first approximation, program compilation should therefore be straightforward. Unfortunately, experience with many different compilers has been uniform: the options available for obtaining optimized code actually produce erroneous code (e.g. operations inside {\tt DO} loops are moved out before them, where some of the variables have not yet been properly set). Therefore the general advice is to use a low optimization level, such as {\tt OPTIMIZE(2)} on IBM, {\tt /NOOPT} on VAX, etc. Note that this is often not the default setting. Specifically on the Apollo, there seems to exist a problem of non-compliance to the Fortran 77 standard for the handling of unformatted write and backspace, as used by {\tt RLUGET} and {\tt RLUSET}. These routines do not form a necessary part of the event generator itself, and so {\tt RLUGET} and {\tt RLUSET} could well just be commented out, or some alternative solution found. {\tt SAVE} statements have been included in accordance with the Fortran standard. Since most ordinary machines take {\tt SAVE} for granted, this part is not particularly well tried out, however. All default settings and particle and process data are stored in {\tt BLOCK DATA LUDATA} for {\Je} and {\tt BLOCK DATA PYDATA} for {\Py}. These subprograms must be linked for a proper functioning of the other routines. On some machines this is not done automatically but must be forced by you, in particular if {\Je} and {\Py} are maintained as libraries from which routines are to be loaded only when they are needed. In this connection we note that the library approach does not give any significant space advantages over a loading of the packages as a whole, since a normal run will call on most of the routines anyway, directly or indirectly. Since most machines in current use are 32-bit ones, this is the precision normally assumed. A few pieces of code have therefore had to be written in double precision. As a rule of thumb, double-precision variables have as first character {\tt D}, but there are a few exceptions. For applications at very high energies, such as LHC, the use of single precision for any real variable is a problem. It might then be necessary to rewrite the program completely, i.e. to have a declaration {\tt IMPLICIT DOUBLE PRECISION(A-H,O-Z)} at the beginning of each subprogram, and to change all real constants to double precision. Needless to say, the latter is a major undertaking. In some cases, shortcuts are available. On the IBM, for instance, the {\tt AUTODBL} compiler option for automatic precision doubling works fine, provided only that an even number of integers precede real numbers in common blocks. In {\Je} you therefore need to introduce an additional integer variable ({\tt NPAD}, say) directly after {\tt N} in the {\tt LUJETS} common block, and in {\Py} an additional integer ({\tt MSEPAD}) after {\tt MSEL} in the {\tt PYSUBS} common block. Some pieces of code will then actually run in quadruple precision. A test program, {\tt LUTEST}\label{p:LUTEST}, is included in the {\Je} package. It is disguised as a subroutine, so you have to run a main program \begin{verbatim} CALL LUTEST(1) END \end{verbatim} This program will generate six hundred events of different types, under a variety of conditions. If {\Je} has not been properly installed, this program is likely to crash, or at least generate a number of erroneous events. This will then clearly be marked in the output, which otherwise will just contain a few sample event listings and a table of the number of different particles produced. To switch off the output of normal events and final table, use {\tt LUTEST(0)} instead of {\tt LUTEST(1)}. The final tally of errors detected should read 0. In exactly the same vein, a test program {\tt PYTEST}\label{p:PYTEST} comes with the {\Py} package. You then have to run a program \begin{verbatim} CALL PYTEST(1) END \end{verbatim} As before the alternative {\tt PYTEST(0)} will give a less extensive listing. No errors should appear during execution. \subsection{Program Philosophy} The Monte Carlo programs are built as slave systems, i.e. you, the user, have to supply the main program. From this the various subroutines are called on to execute specific tasks, after which control is returned to the main program. Some of these tasks may be very trivial, whereas the `high-level' routines by themselves may make a large number of subroutine calls. Many routines are not intended to be called directly by you, but only from higher-level routines such as {\tt LUEXEC}, {\tt LUEEVT}, {\tt PYINIT} or {\tt PYEVNT}. Basically, this means that there are three ways by which you communicate with the programs. First, by setting common block variables, you specify the details of how the programs should perform specific tasks, i.e. which subprocesses should be generated (for {\Py}), which particle masses should be assumed, which coupling constants used, which fragmentation scenarios, and so on with hundreds of options and parameters. Second, by calling subroutines you tell the programs to generate events according to the rules established above. Normally there are few subroutine arguments, and those are usually related to details of the physical situation, such as what c.m. energy to assume for events. Third, you can either look at the common block {\tt LUJETS} to extract information on the generated event, or you can call on various functions and subroutines to analyse the event further for you. It should be noted that, while the physics content is obviously at the centre of attention, the {\JePy} package also contains a very extensive setup of auxiliary service routines. The hope is that this will provide a comfortable working environment, where not only events are generated, but where you also linger on to perform a lot of the subsequent studies. Of course, for detailed studies, it may be necessary to interface the output directly to a detector simulation program. The general rule is that all routines have names that are six characters long, beginning with {\tt LU} for {\Je} routines and {\tt PY} for {\Py} ones. Real-valued functions in {\Je} begin with {\tt UL} instead. There are three exceptions to both the length and the initial character rules: {\tt KLU}, {\tt PLU} and {\tt RLU}. The former two functions are strongly coupled to the {\tt K} and {\tt P} matrices in the {\tt LUJETS} common block, the latter uses {\tt R} to emphasize the r\^ole as a random-number generator. Also common block names are six characters long and start with {\tt LU} or {\tt PY}. On the issue of initialization, {\Je} and {\Py} behave quite differently. Most {\Je} routines work without any initialization (except for the one implied by the presence of {\tt BLOCK DATA LUDATA}, see above), i.e. each event and each task stand on their own. Current common block values are used to perform the tasks in specific ways, and those rules can be changed from one event to the next (or even within the generation of one and the same event) without any penalty. The random-number generator is initialized at the first call, but usually this is transparent. Therefore the two {\Je} routines {\tt LUEEVT} (and some of the routines called by it) and {\tt LUONIA} are basically the only ones to contain some elements of initialization, where there are a few advantages if events are generated in a coherent fashion, but even here the penalty for not doing it is small. In {\Py}, on the other hand, a sizeable amount of initialization is performed in the {\tt PYINIT} call, and thereafter the events generated by {\tt PYEVNT} all obey the rules established at that point. Therefore common block variables that specify methods to be used have to be set before the {\tt PYINIT} call and then not be changed afterwards, with few exceptions. Of course, it is possible to perform several {\tt PYINIT} calls in the same run, but there is a significant time overhead involved, so this is not something one would do for each new event. Apart from writing a title page, giving a brief initialization information, printing error messages if need be, and responding to explicit requests for listings, all tasks of the programs are performed `silently'. All output is directed to unit {\tt MSTU(11)}, by default 6, and it is up to you to set this unit open for write. The only exceptions are {\tt RLUGET}, {\tt RLUSET} and {\tt LUUPDA} where, for obvious reasons, the input/output file number is specified at each call. Here you again have to see to it that proper read/write access is set. The programs are extremely versatile, but the price to be paid for this is having a large number of adjustable parameters and switches for alternative modes of operation. No single user is ever likely to need more than a fraction of the available options. Since all these parameters and switches are assigned sensible default values, there is no reason to worry about them until the need arises. Unless explicitly stated (or obvious from the context) all switches and parameters can be changed independently of each other. One should note, however, that if only a few switches/parameters are changed, this may result in an artificially bad agreement with data. Many disagreements can often be cured by a subsequent retuning of some other parameters of the model, in particular those that were once determined by a comparison with data in the context of the default scenario. For example, for $\ee$ annihilation, such a retuning could involve one QCD parameter ($\alphas$ or $\Lambda$), the longitudinal fragmentation function, and the average transverse fragmentation momentum. The programs contain a number of checks that requested processes have been implemented, that flavours specified for jet systems make sense, that the energy is sufficient to allow hadronization, that the memory space in {\tt LUJETS} is large enough, etc. If anything goes wrong that the program can catch (obviously this may not always be possible), an error message will be printed and the treatment of the corresponding event will be cut short. In serious cases, the program will abort. As long as no error messages appear on the output, it may not be worthwhile to look into the rules for error checking, but if but one message appears, it should be enough cause for alarm to receive prompt attention. Also warnings are sometimes printed. These are less serious, and the experienced user might deliberately do operations which go against the rules, but still can be made to make sense in their context. Only the first few warnings will be printed, thereafter the program will be quiet. By default, the program is set to stop execution after ten errors, after printing the last erroneous event. It must be emphasized that not all errors will be caught. In particular, one tricky question is what happens if an integer-valued common block switch or subroutine/function argument is used with a value that is not defined. In some subroutine calls, a prompt return will be expedited, but in most instances the subsequent action is entirely unpredictable, and often completely haywire. The same goes for real-valued variables that are assigned values outside the physically sensible range. One example will suffice here: if {\tt PARJ(2)} is defined as the $\s / \u$ suppression factor, a value $>1$ will not give more profuse production of $\s$ than of $\u$, but actually a spillover into $\c$ production. Users, beware! \subsection{Manual Conventions} In the manual parts of this report, some conventions are used. All names of subprograms, common blocks and variables are given in upper-case `typewriter' style, e.g. {\tt MSTP(111)=0}. Also program examples are given in this style. If a common block variable must have a value set at the beginning of execution, then a default value is stored in one of the block data subprograms {\tt LUDATA} and {\tt PYDATA}. Such a default value is usually indicated by a `(D=\ldots)' immediately after the variable name, e.g. \begin{entry} \iteme{MSTJ(1) :} (D=1) choice of fragmentation scheme. \end{entry} All variables in the {\Je} common blocks (with very few exceptions, clearly marked) can be freely changed from one event to the next, or even within the treatment of one single event. In the {\Py} common blocks the situation is more complicated. The values of many switches and parameters are used already in the {\tt PYINIT} call, and cannot be changed after that. The problem is mentioned in the preamble to the afflicted common blocks, which in particular means {\tt /PYPARS/} and {\tt /PYSUBS/}. For the variables which may still be changed from one event to the next, a `(C)' is added after the `(D=\ldots)' statement. Normally, variables internal to the program are kept in separate common blocks and arrays, but in a few cases such internal variables appear among arrays of switches and parameters, mainly for historical reasons. These are denoted by `(R)' for variables you may want to read, because they contain potentially interesting information, and by `(I)' for purely internal variables. In neither case may the variables be changed by you. In the description of a switch, the alternatives that this switch may take are often enumerated, e.g. \begin{entry} \iteme{MSTJ(1) :} (D=1) choice of fragmentation scheme. \begin{subentry} \iteme{= 0 :} no jet fragmentation at all. \iteme{= 1 :} string fragmentation according to the Lund model. \iteme{= 2 :} independent fragmentation, according to specification in {\tt MSTJ(2)} and {\tt MSTJ(3)}. \end{subentry} \end{entry} If you then use any value other than 0, 1 or 2, results are unpredictable. The action could even be different in different parts of the program, depending on the order in which the alternatives are identified. It is also up to you to choose physically sensible values for parameters: there is no check on the allowed ranges of variables. We gave an example of this at the end of the preceding section. Subroutines you are expected to use are enclosed in a box at the point where they are defined: \drawbox{CALL LULIST(MLIST)} \boxsep This is followed by a description of input or output parameters. The difference between input and output is not explicitly marked, but should be obvious from the context. In fact, the event-analysis routines of section \ref{ss:evanrout} return values, while all the rest only have input variables. Routines that are only used internally are not boxed in. However, we use boxes for all common blocks, so as to enhance the readability. \subsection{Getting Started with JETSET} \label{ss:JETstarted} As a first example, assume that you want to study the production of $\u \ubar$ 2-jet systems at 20 GeV energy. To do this, write a main program \begin{verbatim} CALL LU2ENT(0,2,-2,20.) CALL LULIST(1) END \end{verbatim} and run this program, linked together with {\Je}. The routine {\tt LU2ENT} is specifically intended for storing two entries (jets or particles). The first argument (0) is a command to perform fragmentation and decay directly after the entries have been stored, the second and third that the two entries are $\u$ (2) and $\ubar$ ($-2$), and the last that the c.m. energy of the pair is 20 GeV. When this is run, the resulting event is stored in the {\tt LUJETS} common block. This information can then be read out by you. No output is produced by {\tt LU2ENT} itself, except for a title page which appears once for every {\JePy} run. Instead the second command, to {\tt LULIST}, provides a simple visible summary of the information stored in {\tt LUJETS}. The argument (1) indicates that the short version should be used, which is suitable for viewing the listing directly on an 80-column terminal screen. It might look as shown here. \begin{verbatim} Event listing (summary) I particle/jet KS KF orig p_x p_y p_z E m 1 (u) A 12 2 0 0.000 0.000 10.000 10.000 0.006 2 (u~) V 11 -2 0 0.000 0.000 -10.000 10.000 0.006 3 (string) 11 92 1 0.000 0.000 0.000 20.000 20.000 4 (rho+) 11 213 3 0.098 -0.154 2.710 2.856 0.885 5 (rho-) 11 -213 3 -0.227 0.145 6.538 6.590 0.781 6 pi+ 1 211 3 0.125 -0.266 0.097 0.339 0.140 7 (Sigma0) 11 3212 3 -0.254 0.034 -1.397 1.855 1.193 8 (K*+) 11 323 3 -0.124 0.709 -2.753 2.968 0.846 9 p~- 1 -2212 3 0.395 -0.614 -3.806 3.988 0.938 10 pi- 1 -211 3 -0.013 0.146 -1.389 1.403 0.140 11 pi+ 1 211 4 0.109 -0.456 2.164 2.218 0.140 12 (pi0) 11 111 4 -0.011 0.301 0.546 0.638 0.135 13 pi- 1 -211 5 0.089 0.343 2.089 2.124 0.140 14 (pi0) 11 111 5 -0.316 -0.197 4.449 4.467 0.135 15 (Lambda0) 11 3122 7 -0.208 0.014 -1.403 1.804 1.116 16 gamma 1 22 7 -0.046 0.020 0.006 0.050 0.000 17 K+ 1 321 8 -0.084 0.299 -2.139 2.217 0.494 18 (pi0) 11 111 8 -0.040 0.410 -0.614 0.751 0.135 19 gamma 1 22 12 0.059 0.146 0.224 0.274 0.000 20 gamma 1 22 12 -0.070 0.155 0.322 0.364 0.000 21 gamma 1 22 14 -0.322 -0.162 4.027 4.043 0.000 22 gamma 1 22 14 0.006 -0.035 0.422 0.423 0.000 23 p+ 1 2212 15 -0.178 0.033 -1.343 1.649 0.938 24 pi- 1 -211 15 -0.030 -0.018 -0.059 0.156 0.140 25 gamma 1 22 18 -0.006 0.384 -0.585 0.699 0.000 26 gamma 1 22 18 -0.034 0.026 -0.029 0.052 0.000 sum: 0.00 0.000 0.000 0.000 20.000 20.000 \end{verbatim} (A few blanks have been removed between the columns to make it fit into the format of this text.) Look in the particle/jet column and note that the first two lines are the original $\u$ and $\ubar$, where `bar' is actually written `$\sim$' to save space in longer names. The parentheses enclosing the names, `{\tt (u)}' and `{\tt (u\~{})}', are there as a reminder that these jets actually have been allowed to fragment. The jets are still retained so that event histories can be studied. Also note that the {\tt KF} (flavour code) column contains 2 in the first line and $-2$ in the second. These are the codes actually stored to denote the presence of a $\u$ and a $\ubar$, cf. the {\tt LU2ENT} call, while the names written are just conveniences used when producing visible output. The {\tt A} and {\tt V} near the end of the particle/jet column indicate the beginning and end of a string (or cluster, or independent fragmentation) parton system; any intermediate entries belonging to the same system would have had an {\tt I} in that column. (This gives a poor man's representation of an up-down arrow, $\updownarrow$.) In the {\tt orig} (origin) column, the zeros indicate that $\u$ and $\ubar$ are two initial entries. The subsequent line, number 3, denotes the fragmenting $\u \ubar$ string system as a whole, and has origin 1, since the first parton of this string system is entry number 1. The particles in lines 4--10 have origin 3 to denote that they come directly from the fragmentation of this string. In string fragmentation it is not meaningful to say that a particle comes from only the $\u$ quark or only the $\ubar$ one. It is the string system as a whole that gives a $\rho^+$, a $\rho^-$, a $\pi^+$, a $\Sigma^0$, a $\K^{*+}$, a $\pbar^-$, and a $\pi^-$. Note that some of the particle names are again enclosed in parentheses, indicating that these particles are not present in the final state either, but have decayed further. Thus the $\pi^-$ in line 13 and the $\pi^0$ in line 14 have origin 5, as an indication that they come from the decay of the $\rho^-$ in line 5. Only the names not enclosed in parentheses remain at the end of the fragmentation/decay chain, and are thus experimentally observable. The actual status code used to distinguish between different classes of entries is given in the {\tt KS} column; codes in the range 1--10 correspond to remaining entries, and those above 10 to those that have fragmented or decayed. The columns with {\tt p\_x}, {\tt p\_y}, {\tt p\_z}, {\tt E} and {\tt m} are quite self-explanatory. All momenta, energies and masses are given in units of GeV, since the speed of light is taken to be $c = 1$. Note that energy and momentum are conserved at each step of the fragmentation/decay process (although there exist options where this is not true). Also note that the $z$ axis plays the r\^ole of preferred direction, along which the original partons are placed. The final line is intended as a quick check that nothing funny happened. It contains the summed charge, summed momentum, summed energy and invariant mass of the final entries at the end of the fragmentation/decay chain, and the values should agree with the input implied by the {\tt LU2ENT} arguments. (In fact, warnings would normally appear on the output if anything untoward happened, but that is another story.) The above example has illustrated roughly what information is to be had in the event record, but not so much about how it is stored. This is better seen by using a 132-column format for listing events. Try e.g. the following program \begin{verbatim} CALL LU3ENT(0,1,21,-1,30.,0.9,0.7) CALL LULIST(2) CALL LUEDIT(3) CALL LULIST(2) END \end{verbatim} where a 3-jet $\d \g \dbar$ event is generated in the first line and listed in the second. This listing will contain the numbers as directly stored in the common block {\tt LUJETS} \begin{verbatim} COMMON/LUJETS/N,K(4000,5),P(4000,5),V(4000,5) \end{verbatim} For particle {\tt I}, {\tt K(I,1)} thus gives information on whether or not a jet or particle has fragmented or decayed, {\tt K(I,2)} gives the particle code, {\tt K(I,3)} its origin, {\tt K(I,4)} and {\tt K(I,5)} the position of fragmentation/decay products, and {\tt P(I,1})--{\tt P(I,5)} momentum, energy and mass. The number of lines in current use is given by {\tt N}, i.e. 1 $\leq$ {\tt I} $\leq$ {\tt N}. The {\tt V} matrix contains decay vertices; to view those {\tt LULIST(3)} has to be used. It is important to learn the rules for how information is stored in {\tt LUJETS}. The third line in the program illustrates another important point about {\Je}: a number of routines are available for manipulating the event record after the event has been generated. Thus {\tt LUEDIT(3)} will remove everything except stable charged particles, as shown by the result of the second {\tt LULIST} call. More advanced possibilities include things like sphericity or clustering routines. Apart from the input arguments of subroutine calls, control on the doings of {\Je} may be imposed via the {\tt LUDAT1}, {\tt LUDAT2}, {\tt LUDAT3} and {\tt LUDAT4} common blocks. Here sensible default values are always provided. A user might want to switch off all particle decays by putting {\tt MSTJ(21)=0} or increase the $\s / \u$ ratio in fragmentation by putting {\tt PARJ(2)=0.40}, to give but two examples. It is by exploring the possibilities offered here that {\Je} can be turned into an extremely versatile tool, even if all the nice physics is already present in the default values. As a final, semirealistic example, assume that the $\pT$ spectrum of $\pi^+$ particles is to be studied in 91.2 GeV $\ee$ annihilation events, where $\pT$ is to be defined with respect to the sphericity axis. Using the HBOOK package (version 4, watch out for version- or installation-specific differences) for histogramming, a complete program might look like \begin{verbatim} C...Common blocks. COMMON/LUJETS/N,K(4000,5),P(4000,5),V(4000,5) COMMON/PAWC/HMEMOR(10000) C...Reserve histogram memory and book histograms. CALL HLIMIT(10000) CALL HBOOK1(1,'pT spectrum of pi+',100,0.,5.,0.) C...Number of events to generate. Loop over events. NEVT=100 DO 110 IEVT=1,NEVT C...Generate event. List first one. CALL LUEEVT(0,91.2) IF(IEVT.EQ.1) CALL LULIST(1) C...Find sphericity axis and rotate event so sphericity along z axis. CALL LUSPHE(SPH,APL) CALL LUEDIT(31) C...Loop over all particles, but skip if not pi+. DO 100 I=1,N IF(K(I,2).NE.211) GOTO 100 C...Calculate pT and fill in histogram. PT=SQRT(P(I,1)**2+P(I,2)**2) CALL HF1(1,PT,1.) C...End of particle and event loops. 100 CONTINUE 110 CONTINUE C...Normalize histogram properly and list it. CALL HOPERA(1,'+',1,1,20./NEVT,0.) CALL HISTDO END \end{verbatim} Study this program, try to understand what happens at each step, and run it to check that it works. You should then be ready to look at the relevant sections of this report and start writing your own programs. \subsection{Getting Started with PYTHIA} \label{ss:PYTstarted} A {\Py} run has to be more strictly organized than a {\Je} one, in that it is necessary to initialize the generation before events can be generated, and in that it is not possible to change switches and parameters freely during the course of the run. A fairly precise recipe for how a run should be structured can therefore be given. Thus, the usage of {\Py} can be subdivided into three steps. \begin{Enumerate} \item The initialization step. It is here that all the basic characteristics of the coming generation are specified. The material in this section includes the following. \begin{Itemize} \item Common blocks, at least the following, and maybe some more: \begin{verbatim} COMMON/LUJETS/N,K(4000,5),P(4000,5),V(4000,5) COMMON/LUDAT1/MSTU(200),PARU(200),MSTJ(200),PARJ(200) COMMON/PYSUBS/MSEL,MSUB(200),KFIN(2,-40:40),CKIN(200) COMMON/PYPARS/MSTP(200),PARP(200),MSTI(200),PARI(200) \end{verbatim} \vspace{-\baselineskip} \item Selection of required processes. Some fixed `menus' of subprocesses can be selected with different {\tt MSEL} values, but with {\tt MSEL}=0 it is possible to compose `\`a la carte', using the subprocess numbers. To generate processes 14, 18 and 29, for instance, one needs \begin{verbatim} MSEL=0 MSUB(14)=1 MSUB(18)=1 MSUB(29)=1 \end{verbatim} \vspace{-\baselineskip} \item Selection of kinematics cuts in the {\tt CKIN} array. To generate hard scatterings with 5~GeV $\leq \pT \leq$ 10~GeV, for instance, use \begin{verbatim} CKIN(3)=5. CKIN(4)=10. \end{verbatim} \vspace{-\baselineskip} Unfortunately, initial- and final-state radiation will shift around the kinematics of the hard scattering, making the effects of cuts less predictable. One therefore always has to be very careful that no desired event configurations are cut out. \item Definition of underlying physics scenario, e.g. top mass. \item Selection of parton-distribution sets, $Q^2$ definitions, and all other details of the generation. \item Switching off of generator parts not needed for toy simulations, e.g. fragmentation for parton level studies. \item Initialization of the event generation procedure. Here kinematics is set up, maxima of differential cross sections are found for future Monte Carlo generation, and a number of other preparatory tasks carried out. Initialization is performed by {\tt PYINIT}, which should be called only after the switches and parameters above have been set to their desired values. The frame, the beam particles and the energy have to be specified. \begin{verbatim} CALL PYINIT('CMS','p','pbar',1800.) \end{verbatim} \vspace{-\baselineskip} \item Any other initial material required by the user, e.g. histogram booking. \end{Itemize} \item The generation loop. It is here that events are generated and studied. It includes the following tasks: \begin{Itemize} \item Generation of the next event, with \begin{verbatim} CALL PYEVNT \end{verbatim} \vspace{-\baselineskip} \item Printing of a few events, to check that everything is working as planned, with \begin{verbatim} CALL LULIST(1) \end{verbatim} \vspace{-\baselineskip} \item An analysis of the event for properties of interest, either directly reading out information from the {\tt LUJETS} common block or making use of a number of utility routines in {\Je}. \item Saving of events on tape, or interfacing to detector simulation. \end{Itemize} \item The finishing step. Here the tasks are: \begin{Itemize} \item Printing a table of deduced cross sections, obtained as a by-product of the Monte Carlo generation activity, with the command \begin{verbatim} CALL PYSTAT(1) \end{verbatim} \vspace{-\baselineskip} \item Printing histograms and other user output. \end{Itemize} \end{Enumerate} To illustrate this structure, imagine a toy example, where one wants to simulate the production of a 300 GeV Higgs particle. In {\Py}, a program for this might look something like the following. \begin{verbatim} C...Common blocks. COMMON/LUJETS/N,K(4000,5),P(4000,5),V(4000,5) COMMON/LUDAT1/MSTU(200),PARU(200),MSTJ(200),PARJ(200) COMMON/LUDAT2/KCHG(500,3),PMAS(500,4),PARF(2000),VCKM(4,4) COMMON/LUDAT3/MDCY(500,3),MDME(2000,2),BRAT(2000),KFDP(2000,5) COMMON/PYSUBS/MSEL,MSUB(200),KFIN(2,-40:40),CKIN(200) COMMON/PYPARS/MSTP(200),PARP(200),MSTI(200),PARI(200) COMMON/PAWC/HBOOK(10000) C...Number of events to generate. Switch on proper processes. NEV=1000 MSEL=0 MSUB(102)=1 MSUB(123)=1 MSUB(124)=1 C...Select t and H masses and kinematics cuts in mass. PMAS(6,1)=140. PMAS(25,1)=300. CKIN(1)=290. CKIN(2)=310. C...For simulation of hard process only: cut out unnecessary tasks. MSTP(61)=0 MSTP(71)=0 MSTP(81)=0 MSTP(111)=0 C...Initialize and list partial widths. CALL PYINIT('CMS','p','p',16000.) CALL PYSTAT(2) C...Book histograms. CALL HLIMIT(10000) CALL HBOOK1(1,'Higgs mass',50,275.,325.,0.) C...Generate events. Look at first few. DO 200 IEV=1,NEV CALL PYEVNT IF(IEV.LE.3) CALL LULIST(1) C...Loop over particles to find Higgs and histogram its mass. DO 100 I=1,N 100 IF(K(I,2).EQ.25) HMASS=P(I,5) CALL HF1(1,HMASS,1.) 200 CONTINUE C...Print cross sections and histograms. CALL PYSTAT(1) CALL HISTDO END \end{verbatim} Here 102, 123 and 124 are the three main Higgs production graphs $\g \g \rightarrow \H$, $\Z \Z \rightarrow \H$, and $\W \W \rightarrow \H$, and {\tt MSUB(ISUB)=1} is the command to switch on process {\tt ISUB}. Full freedom to combine subprocesses `\`a la carte' is ensured by {\tt MSEL=0}; ready-made `menus' can be ordered with other {\tt MSEL} numbers. The {\tt PMAS} commands set the masses of the top quark and the Higgs itself, and the {\tt CKIN} variables the desired mass range of the Higgs --- a Higgs with a 300 GeV nominal mass actually has a fairly broad Breit--Wigner type mass distribution. The {\tt MSTP} switches that come next are there to modify the generation procedure, in this case to switch off initial- and final-state radiation, multiple interactions among beam jets, and fragmentation, to give only the `parton skeleton' of the hard process. The {\tt PYINIT} call initializes {\Py}, by finding maxima of cross sections, recalculating the Higgs decay properties (which depend on the Higgs mass), etc. The decay properties can be listed with {\tt PYSTAT(2)}. Inside the event loop, {\tt PYEVNT} is called to generate an event, and {\tt LULIST(1)} to list the event. The information used by {\tt LULIST(1)} is the event record, stored in the common block {\tt LUJETS}. Here one finds all produced particles, both final and intermediate ones, with information on particle species and event history ({\tt K} array), particle momenta ({\tt P} array) and production vertices ({\tt V} array). In the loop over all particles produced, {\tt 1} through {\tt N}, the Higgs particle is found by its code, {\tt K(I,2)=25}, and its mass is stored in {\tt P(I,5)}. After all events have been generated, {\tt PYSTAT(1)} gives a summary of the number of events generated in the various allowed channels, and the inferred cross sections. In the run above, a typical event listing might look like the following. \begin{verbatim} Event listing (summary) I particle/jet KF p_x p_y p_z E m 1 !p+! 2212 0.000 0.000 8000.000 8000.000 0.938 2 !p+! 2212 0.000 0.000-8000.000 8000.000 0.938 ====================================================================== 3 !g! 21 -0.505 -0.229 28.553 28.558 0.000 4 !g! 21 0.224 0.041 -788.073 788.073 0.000 5 !g! 21 -0.505 -0.229 28.553 28.558 0.000 6 !g! 21 0.224 0.041 -788.073 788.073 0.000 7 !H0! 25 -0.281 -0.188 -759.520 816.631 300.027 8 !W+! 24 120.648 35.239 -397.843 424.829 80.023 9 !W-! -24 -120.929 -35.426 -361.677 391.801 82.579 10 !e+! -11 12.922 -4.760 -160.940 161.528 0.001 11 !nu_e! 12 107.726 39.999 -236.903 263.302 0.000 12 !s! 3 -62.423 7.195 -256.713 264.292 0.199 13 !c~! -4 -58.506 -42.621 -104.963 127.509 1.350 ====================================================================== 14 (H0) 25 -0.281 -0.188 -759.520 816.631 300.027 15 (W+) 24 120.648 35.239 -397.843 424.829 80.023 16 (W-) -24 -120.929 -35.426 -361.677 391.801 82.579 17 e+ -11 12.922 -4.760 -160.940 161.528 0.001 18 nu_e 12 107.726 39.999 -236.903 263.302 0.000 19 s A 3 -62.423 7.195 -256.713 264.292 0.199 20 c~ V -4 -58.506 -42.621 -104.963 127.509 1.350 21 ud_1 A 2103 -0.101 0.176 7971.328 7971.328 0.771 22 d V 1 -0.316 0.001 -87.390 87.390 0.010 23 u A 2 0.606 0.052 -0.751 0.967 0.006 24 uu_1 V 2203 0.092 -0.042-7123.668 7123.668 0.771 ====================================================================== sum: 2.00 0.00 0.00 0.00 15999.98 15999.98 \end{verbatim} The above event listing is abnormally short, in part because some columns of information were removed to make it fit into this text, in part because all initial- and final-state QCD radiation, all non-trivial beam jet structure, and all fragmentation was inhibited in the generation. Therefore only the skeleton of the process is visible. In lines 1 and 2 one recognizes the two incoming protons. In lines 3 and 4 are incoming partons before initial-state radiation and in 5 and 6 after --- since there is no such radiation they coincide here. Line 7 shows the Higgs produced by $\g \g$ fusion, 8 and 9 its decay products and 10--13 the second-step decay products. Up to this point lines give a summary of the event history, indicated by the exclamation marks that surround particle names (and also reflected in the {\tt K(I,1)} code, not shown). From line 14 onwards come the particles actually produced in the final states, first in lines 14--16 particles that subsequently decayed, which have their names surrounded by brackets, and finally the particles and jets left in the end, including beam remnants. Here this also includes a number of unfragmented jets, since fragmentation was inhibited. Ordinarily, the listing would have gone on for a few hundred more lines, with the particles produced in the fragmentation and their decay products. The final line gives total charge and momentum, as a convenient check that nothing unexpected happened. The first column of the listing is just a counter, the second gives the particle name and information on status and string drawing (the {\tt A} and {\tt V}), the third the particle-flavour code (which is used to give the name), and the subsequent columns give the momentum components. One of the main problems is to select kinematics efficiently. Imagine for instance that one is interested in the production of a single $\Z$ with a transverse momentum in excess of 50 GeV. If one tries to generate the inclusive sample of $\Z$ events, by the basic production graphs $\q \qbar \rightarrow \Z$, then most events will have low transverse momenta and will have to be discarded. That any of the desired events are produced at all is due to the initial-state generation machinery, which can build up transverse momenta for the incoming $\q$ and $\qbar$. However, the amount of initial-state radiation cannot be constrained beforehand. To increase the efficiency, one may therefore turn to the higher-order processes $\q \g \rightarrow \Z \q$ and $\q \qbar \rightarrow \Z \g$, where already the hard subprocess gives a transverse momentum to the $\Z$. This transverse momentum can be constrained as one wishes, but again initial- and final-state radiation will smear the picture. If one were to set a $\pT$ cut at 50 GeV for the hard-process generation, those events where the $\Z$ was given only 40 GeV in the hard process but got the rest from initial-state radiation would be missed. Not only therefore would cross sections come out wrong, but so might the typical event shapes. In the end, it is therefore necessary to find some reasonable compromise, by starting the generation at 30 GeV, say, if one knows that only rarely do events below this value fluctuate up to 50 GeV. Of course, most events will therefore not contain a $\Z$ above 50 GeV, and one will have to live with some inefficiency. It is not uncommon that only one event out of ten can be used, and occasionally it can be even worse. If it is difficult to set kinematics, it is often easier to set the flavour content of a process. In a Higgs study, one might wish, for example, to consider the decay $\H^0 \rightarrow \Z^0 \Z^0$, with each $\Z^0 \rightarrow \ee$ or $\mu^+ \mu^-$. It is therefore necessary to inhibit all other $\H^0$ and $\Z^0$ decay channels, and also to adjust cross sections to take into account this change, all of which is fairly straightforward. However, if one wanted to consider instead the decay $\Z^0 \rightarrow \c \cbar$, with a $\D$ meson producing a lepton, not only would there then be the problem of different leptonic branching ratios for different $\D$:s (which means that fragmentation and decay treatments would no longer decouple), but also that of additional $\c \cbar$ pair production in parton-shower evolution, at a rate that is unknown beforehand. In practice, it is therefore impossible to force $\D$ decay modes in a consistent manner. \clearpage \section{Monte Carlo Techniques} Quantum mechanics introduces a concept of randomness in the behaviour of physical processes. The virtue of event generators is that this randomness can be simulated by the use of Monte Carlo techniques. In the process, the program authors have to use some ingenuity to find the most efficient way to simulate an assumed probability distribution. A detailed description of possible techniques would carry us too far, but in this section some of the most frequently used approaches are presented, since they will appear in discussions in subsequent sections. Further examples may be found e.g. in \cite{Jam80}. First of all one assumes the existence of a random number generator. This is a (Fortran) function which, each time it is called, returns a number $R$ in the range between 0 and 1, such that the inclusive distribution of numbers $R$ is flat in the range, and such that different numbers $R$ are uncorrelated. The random number generator that comes with {\Je} is described at the end of this section, and we defer the discussion until then. \subsection{Selection From a Distribution} \label{ss:MCdistsel} The situation that is probably most common is that we know a function $f(x)$ which is non-negative in the allowed $x$ range $x_{\mmin} \leq x \leq x_{\mmax}$. We want to select an $x$ `at random' so that the probability for a given $x$ is proportional to $f(x)$. Here $f(x)$ might be a fragmentation function, a differential cross section, or any of a number of distributions. One does not have to assume that the integral of $f(x)$ is explicitly normalized to unity: by the Monte Carlo procedure of picking exactly one accepted $x$ value, normalization is implicit in the final result. Sometimes the integral of $f(x)$ does carry a physics content of its own, as part of an overall weight factor we want to keep track of. Consider, for instance, the case when $x$ represents one or several phase-space variables and $f(x)$ a differential cross section; here the integral has a meaning of total cross section for the process studied. The task of a Monte Carlo is then, on the one hand, to generate events one at a time, and, on the other hand, to estimate the total cross section. The discussion of this important example is deferred to section \ref{ss:PYTcrosscalc}. If it is possible to find a primitive function $F(x)$ which has a known inverse $F^{-1}(x)$, an $x$ can be found as follows (method 1): \begin{eqnarray} & \displaystyle{ \int_{x_{\mmin}}^{x} f(x) \, \d x = R \int_{x_{\mmin}}^{x_{\mmax}} f(x) \, \d x } & \nonumber \\[1mm] \Longrightarrow & x = F^{-1}(F(x_{\mmin}) + R(F(x_{\mmax}) - F(x_{\mmin}))) ~. & \end{eqnarray} The statement of the first line is that a fraction $R$ of the total area under $f(x)$ should be to the left of $x$. However, seldom are functions of interest so nice that the method above works. It is therefore necessary to use more complicated schemes. Special tricks can sometimes be found. Consider e.g. the generation of a Gaussian $f(x) = \exp(-x^2)$. This function is not integrable, but if we combine it with the same Gaussian distribution of a second variable $y$, it is possible to transform to polar coordinates \begin{equation} f(x) \, \d x \, f(y) \, \d y = \exp(-x^2-y^2) \, \d x \, \d y = r \exp(-r^2) \, \d r \, \d \varphi ~, \end{equation} and now the $r$ and $\varphi$ distributions may be easily generated and recombined to yield $x$. At the same time we get a second number $y$, which can also be used. For the generation of transverse momenta in fragmentation, this is very convenient, since in fact we want to assign two transverse degrees of freedom. If the maximum of $f(x)$ is known, $f(x) \leq f_{\mmax}$ in the $x$ range considered, a hit-or-miss method will always yield the correct answer (method 2): \begin{Enumerate} \item select an $x$ with even probability in the allowed range, i.e. $x = x_{\mmin} + R(x_{\mmax} - x_{\mmin})$; \item compare a (new) $R$ with the ratio $f(x)/f_{\mmax}$; if $f(x)/f_{\mmax} \le R$, then reject the $x$ value and return to point 1 for a new try; \item otherwise the most recent $x$ value is retained as final answer. \end{Enumerate} The probability that $f(x)/f_{\mmax} > R$ is proportional to $f(x)$; hence the correct distribution of retained $x$ values. The efficiency of this method, i.e. the average probability that an $x$ will be retained, is $(\int \, f(x) \, \d x) / (f_{\mmax}(x_{\mmax} - x_{\mmin}))$. The method is acceptable if this number is not too low, i.e. if $f(x)$ does not fluctuate too wildly. Very often $f(x)$ does have narrow spikes, and it may not even be possible to define an $f_{\mmax}$. An example of the former phenomenon is a function with a singularity just outside the allowed region, an example of the latter an integrable singularity just at the $x_{\mmin}$ and/or $x_{\mmax}$ borders. Variable transformations may then be used to make a function smoother. Thus a function $f(x)$ which blows up as $1/x$ for $x \rightarrow 0$, with an $x_{\mmin}$ close to 0, would instead be roughly constant if transformed to the variable $y = \ln x$. The variable transformation strategy may be seen as a combination of methods 1 and 2, as follows. Assume the existence of a function $g(x)$, with $f(x) \leq g(x)$ over the $x$ range of interest. Here $g(x)$ is picked to be a `simple' function, such that the primitive function $G(x)$ and its inverse $G^{-1}(x)$ are known. Then (method 3): \begin{Enumerate} \item select an $x$ according to the distribution $g(x)$, using method 1; \item compare a (new) $R$ with the ratio $f(x)/g(x)$; if $f(x)/g(x) \le R$, then reject the $x$ value and return to point 1 for a new try; \item otherwise the most recent $x$ value is retained as final answer. \end{Enumerate} This works, since the first step will select $x$ with a probability $g(x) \, \d x = \d G(x)$ and the second retain this choice with probability $f(x)/g(x)$. The total probability to pick a value $x$ is then just the product of the two, i.e. $f(x) \, \d x$. If $f(x)$ has several spikes, method 3 may work for each spike separately, but it may not be possible to find a $g(x)$ that covers all of them at the same time, and which still has an invertible primitive function. However, assume that we can find a function $g(x) = \sum_i g_i(x)$, such that $f(x) \leq g(x)$ over the $x$ range considered, and such that the functions $g_i(x)$ each are non-negative and simple, in the sense that we can find primitive functions and their inverses. In that case (method 4): \begin{Enumerate} \item select an $i$ at random, with relative probability given by the integrals \begin{equation} \int_{x_{\mmin}}^{x_{\mmax}} g_i(x) \, \d x = G_i(x_{\mmax}) - G_i(x_{\mmin}) ~; \nonumber \end{equation} \item for the $i$ selected, use method 1 to find an $x$, i.e. \begin{equation} x = G_i^{-1}(G_i(x_{\mmin}) + R(G_i(x_{\mmax})-G_i(x_{\mmin}))) ~; \nonumber \end{equation} \item compare a (new) $R$ with the ratio $f(x)/g(x)$; if $f(x)/g(x) \le R$, then reject the $x$ value and return to point 1 for a new try; \item otherwise the most recent $x$ value is retained as final answer. \end{Enumerate} This is just a trivial extension of method 3, where steps 1 and 2 ensure that, on the average, each $x$ value picked there is distributed according to $g(x)$: the first step picks $i$ with relative probability $\int g_i(x) \, \d x$, the second $x$ with absolute probability $g_i(x) / \int g_i(x) \, \d x$ (this is one place where one must remember to do normalization correctly); the product of the two is therefore $g_i(x)$ and the sum over all $i$ gives back $g(x)$. We have now arrived at an approach that is sufficiently powerful for a large selection of problems. In general, for a function $f(x)$ which is known to have sharp peaks in a few different places, the generic behaviour at each peak separately may be covered by one or a few simple functions $g_i(x)$, to which one adds a few more $g_i(x)$ to cover the basic behaviour away from the peaks. By a suitable selection of the relative strengths of the different $g_i$'s, it is possible to find a function $g(x)$ that matches well the general behaviour of $f(x)$, and thus achieve a reasonable Monte Carlo efficiency. The major additional complication is when $x$ is a multidimensional variable. Usually the problem is not so much $f(x)$ itself, but rather that the phase-space boundaries may be very complicated. If the boundaries factorize it is possible to pick phase-space points restricted to the desired region. Otherwise the region may have to be inscribed in a hyper-rectangle, with points picked within the whole hyper-rectangle but only retained if they are inside the allowed region. This may lead to a significant loss in efficiency. Variable transformations may often make the allowed region easier to handle. There are two main methods to handle several dimensions, each with its set of variations. The first method is based on a factorized ansatz, i.e. one attempts to find a function $g(\mbf{x})$ which is everywhere larger than $f(\mbf{x})$, and which can be factorized into $g(\mbf{x}) = g^{(1)}(x_1) \, g^{(2)}(x_2) \cdots g^{(n)}(x_n)$, where $\mbf{x} = (x_1,x_2,\ldots,x_n)$. Here each $g^{(j)}(x_j)$ may in its turn be a sum of functions $g^{(j)}_i$, as in method 4 above. First, each $x_j$ is selected independently, and afterwards the ratio $f(\mbf{x})/g(\mbf{x})$ is used to determine whether to retain the point. The second method is useful if the boundaries of the allowed region can be written in a form where the maximum range of $x_1$ is known, the allowed range of $x_2$ only depends on $x_1$, that of $x_3$ only on $x_1$ and $x_2$, and so on until $x_n$, whose range may depend on all the preceding variables. In that case it may be possible to find a function $g(\mbf{x})$ that can be integrated over $x_2$ through $x_n$ to yield a simple function of $x_1$, according to which $x_1$ is selected. Having done that, $x_2$ is selected according to a distribution which now depends on $x_1$, but with $x_3$ through $x_n$ integrated over. In particular, the allowed range for $x_2$ is known. The procedure is continued until $x_n$ is reached, where now the function depends on all the preceding $x_j$ values. In the end, the ratio $f(\mbf{x})/g(\mbf{x})$ is again used to determine whether to retain the point. \subsection{The Veto Algorithm} \label{ss:vetoalg} The `radioactive decay' type of problems is very common, in particular in parton showers, but it is also used, e.g. in the multiple interactions description in {\Py}. In this kind of problems there is one variable $t$, which may be thought of as giving a kind of time axis along which different events are ordered. The probability that `something will happen' (a nucleus decay, a parton branch) at time $t$ is described by a function $f(t)$, which is non-negative in the range of $t$ values to be studied. However, this na\"{\i}ve probability is modified by the additional requirement that something can only happen at time $t$ if it did not happen at earlier times $t' < t$. (The original nucleus cannot decay once again if it already did decay; possibly the decay products may decay in their turn, but that is another question.) The probability that nothing has happened by time $t$ is expressed by the function ${\cal N}(t)$ and the differential probability that something happens at time $t$ by ${\cal P}(t)$. The basic equation then is \begin{equation} {\cal P}(t) = - \frac{\d {\cal N}}{\d t} = f(t) \, {\cal N}(t) ~. \end{equation} For simplicity, we shall assume that the process starts at time $t = 0$, with ${\cal N}(0) = 1$. The above equation can be solved easily if one notes that $\d {\cal N} / {\cal N} = \d \ln {\cal N}$: \begin{equation} {\cal N}(t) = {\cal N}(0) \exp \left\{ - \int_0^t f(t') \, \d t' \right\} = \exp \left\{ - \int_0^t f(t') \, \d t' \right\} ~, \end{equation} and thus \begin{equation} {\cal P}(t) = f(t) \exp \left\{ - \int_0^t f(t') \, \d t' \right\} ~. \label{mc:Pveto} \end{equation} With $f(t) = c$ this is nothing but the textbook formulae for radioactive decay. In particular, at small times the correct decay probability, ${\cal P}(t)$, agrees well with the input one, $f(t)$, since the exponential factor is close to unity there. At larger $t$, the exponential gives a dampening which ensures that the integral of ${\cal P}(t)$ never can exceed unity, even if the integral of $f(t)$ does. The exponential can be seen as the probability that nothing happens between the original time 0 and the final time $t$. In the parton-shower language, this is (almost) the so-called Sudakov form factor. If $f(t)$ has a primitive function with a known inverse, it is easy to select $t$ values correctly: \begin{equation} \int_0^t {\cal P}(t') \, \d t' = {\cal N}(0) - {\cal N}(t) = 1 - \exp \left\{ - \int_0^t f(t') \, \d t' \right\} = 1 - R ~, \end{equation} which has the solution \begin{equation} F(0) - F(t) = \ln R ~~~ \Longrightarrow ~~~ t = F^{-1}(F(0) - \ln R) ~. \end{equation} If $f(t)$ is not sufficiently nice, one may again try to find a better function $g(t)$, with $f(t) \leq g(t)$ for all $t \geq 0$. However to use method 3 with this $g(t)$ would not work, since the method would not correctly take into account the effects of the exponential term in ${\cal P}(t)$. Instead one may use the so-called veto algorithm: \begin{Enumerate} \item start with $i = 0$ and $t_0 = 0$; \item add 1 to $i$ and select $t_i = G^{-1}(G(t_{i-1}) - \ln R)$, i.e. according to $g(t)$, but with the constraint that $t_i > t_{i-1}$, \item compare a (new) $R$ with the ratio $f(t_i)/g(t_i)$; if $f(t_i)/g(t_i) \le R$, then return to point 2 for a new try; \item otherwise $t_{i}$ is retained as final answer. \end{Enumerate} It may not be apparent why this works. Consider, however, the various ways in which one can select a specific time $t$. The probability that the first try works, $t = t_1$, i.e. that no intermediate $t$ values need be rejected, is given by \begin{equation} {\cal P}_0 (t) = \exp \left\{ - \int_0^t g(t') \, \d t' \right\} \, g(t) \, \frac{f(t)}{g(t)} = f(t) \exp \left\{ - \int_0^t g(t') \, \d t' \right\} ~, \end{equation} where the exponential times $g(t)$ comes from eq. (\ref{mc:Pveto}) applied to $g$, and the ratio $f(t)/g(t)$ is the probability that $t$ is accepted. Now consider the case where one intermediate time $t_1$ is rejected and $t = t_2$ is only accepted in the second step. This gives \begin{equation} {\cal P}_1 (t) = \int_0^t \d t_1 \exp \left\{ - \int_0^{t_1} g(t') \, \d t' \right\} g(t_1) \left[ 1 - \frac{f(t_1)}{g(t_1)} \right] \exp \left\{ - \int_{t_1}^t g(t') \, \d t' \right\} g(t) \, \frac{f(t)}{g(t)} ~, \end{equation} where the first exponential times $g(t_1)$ gives the probability that $t_1$ is first selected, the square brackets the probability that $t_1$ is subsequently rejected, the following piece the probability that $t = t_2$ is selected when starting from $t_1$, and the final factor that $t$ is retained. The whole is to be integrated over all possible intermediate times $t_1$. The exponentials together give an integral over the range from 0 to $t$, just as in ${\cal P}_0$, and the factor for the final step being accepted is also the same, so therefore one finds that \begin{equation} {\cal P}_1 (t) = {\cal P}_0 (t) \int_0^t \d t_1 \left[ g(t_1) - f(t_1) \right] ~. \end{equation} This generalizes. In ${\cal P}_2$ one has to consider two intermediate times, $0 \leq t_1 \leq t_2 \leq t_3 = t$, and so \begin{eqnarray} {\cal P}_2 (t) & = & {\cal P}_0 (t) \int_0^t \d t_1 \left[ g(t_1) - f(t_1) \right] \int_{t_1}^t \d t_2 \left[ g(t_2) - f(t_2) \right] \nonumber \\ & = & {\cal P}_0 (t) \frac{1}{2} \left( \int_0^t \left[ g(t') - f(t') \right] \d t' \right)^2 ~. \end{eqnarray} The last equality is most easily seen if one also considers the alternative region $0 \leq t_2 \leq t_1 \leq t$, where the r\^oles of $t_1$ and $t_2$ have just been interchanged, and the integral therefore has the same value as in the region considered. Adding the two regions, however, the integrals over $t_1$ and $t_2$ decouple, and become equal. In general, for ${\cal P}_i$, the $i$ intermediate times can be ordered in $i!$ different ways. Therefore the total probability to accept $t$, in any step, is \begin{eqnarray} {\cal P} (t) & = & \displaystyle{ \sum_{i=0}^{\infty} {\cal P}_i (t) = {\cal P}_0 (t) \sum_{i=0}^{\infty} \frac{1}{i!} \left( \int_0^t \left[ g(t') - f(t') \right] \d t' \right)^i } \nonumber \\ & = & \displaystyle{ f(t) \exp \left\{ - \int_0^t g(t') \, \d t' \right\} \exp \left\{ \int_0^t \left[ g(t') - f(t') \right] \d t' \right\} } \nonumber \\ & = & \displaystyle{ f(t) \exp \left\{ - \int_0^t f(t') \, \d t' \right\} ~, } \end{eqnarray} which is the desired answer. If the process is to be stopped at some scale $t_{\mmax}$, i.e. if one would like to remain with a fraction ${\cal N}(t_{\mmax})$ of events where nothing happens at all, this is easy to include in the veto algorithm: just iterate upwards in $t$ at usual, but stop the process if no allowed branching is found before $t_{\mmax}$. Usually $f(t)$ is a function also of additional variables $x$. The methods of the preceding subsection are easy to generalize if one can find a suitable function $g(t,x)$ with $f(t,x) \leq g(t,x)$. The $g(t)$ used in the veto algorithm is the integral of $g(t,x)$ over $x$. Each time a $t_i$ has been selected also an $x_i$ is picked, according to $g(t_i,x) \, dx$, and the $(t,x)$ point is accepted with probability $f(t_i,x_i)/g(t_i,x_i)$. \subsection{The Random Number Generator} The construction of a good, portable (pseudo)random generator is not a trivial task. Therefore {\Je} has traditionally stayed away from that area, and just provided the routine {\tt RLU} as an interface, which the user could modify to call on an existing routine, implemented on the actual machine being used. In recent years, progress has been made in constructing portable generators with large periods and other good properties; see the review \cite{Jam90}. Therefore the current version contains a random number generator based on the algorithm proposed by Marsaglia, Zaman and Tsang \cite{Mar90}. This routine should work on any machine with a mantissa of at least 24 digits, i.e. all common 32-bit (or more) computers. Given the same initial state, the sequence will also be identical on different machines. This need not mean that the same sequence of events will be generated on an IBM and a VAX, say, since the different treatments of roundoff errors in numerical operations will lead to slightly different real numbers being tested against these random numbers in IF statements. Also code optimization may lead to a divergence. Apart from nomenclature issues, and the coding of {\tt RLU} as a function rather than a subroutine, the only difference between the {\Je} code and the code given in \cite{Jam90} is that slightly different algorithms are used to ensure that the random number is not equal to 0 or 1 within the machine precision. The generator has a period of over $10^{43}$, and the possibility to obtain almost $10^9$ different and disjoint subsequences, selected by giving an initial integer number. The price to be paid for the long period is that the state of the generator at a given moment cannot be described by a single integer, but requires about 100 words. Some of these are real numbers, and are thus not correctly represented in decimal form. The normal procedure, which makes it possible to restart the generation from a seed value written to the run output, is therefore not convenient. The CERN library implementation keeps track of the number of random numbers generated since the start. With this value saved, in a subsequent run the random generator can be asked to skip ahead the corresponding number of random numbers. {\Je} is a heavy user of random numbers, however: typically 30\% of the full run time is spent on random number generation. Of this, half is overhead coming from the function call administration, but the other half is truly related to the speed of the algorithm. Therefore a skipping ahead would take place with 15\% of the time cost of the original run, i.e. an uncomfortably high figure. Instead a different solution is chosen here. Two special routines are provided for writing and reading the state of the random number generator (plus some initialization information) on a sequential file, in a machine-dependent internal representation. The file used for this purpose has to be specified by you, and opened for read and write. A state is written as a single record, in free format. It is possible to write an arbitrary number of states on a file, and a record can be overwritten, if so desired. The event generation loop might then look something like: \begin{Enumerate} \item save the state of the generator on file (using flag set in point 3 below), \item generate an event, \item study the event for errors or other reasons why to regenerate it later; set flag to overwrite previous generator state if no errors, otherwise set flag to create new record; \item loop back to point 1. \end{Enumerate} With this procedure, the file will contain the state before each of the problematical events. An alternative approach might be to save the state every 100 events or so. If the events are subsequently processed through a detector simulation, you may have to save also other sets of seeds, naturally. In addition to the service routines, the common block which contains the state of the generator is available for manipulation, if you so desire. In particular, the initial seed value is by default 19780503, i.e. different from the Marsaglia/CERN default 54217137. It is possible to change this value before any random numbers have been generated, or to force reinitialization in mid-run with any desired new seed. Inside {\JePy}, some initialization may take place in connection with the very first event generated in a run, so sometimes it may be necessary to generate one ordinary event before reading in a saved state to generate an interesting event. In the current {\Py} version, some of the multiple interaction machinery options contain an element of learning, which means that the event sequence may be broken. It should be noted that, of course, the appearance of a random number generator package inside {\Je} does in no way preclude the use of other routines. You can easily revert to the old approach, where {\tt RLU} is nothing but an interface to an arbitrary external random number generator; e.g. to call a routine {\tt RNDM} all you need to have is \begin{verbatim} FUNCTION RLU(IDUMMY) 100 RLU=RNDM(IDUMMY) IF(RLU.LE.0..OR.RLU.GE.1.) GOTO 100 RETURN END \end{verbatim} The random generator subpackage consists of the following components. \drawbox{R = RLU(IDUMMY)}\label{p:RLU} \begin{entry} \itemc{Purpose:} to generate a (pseudo)random number {\tt R} uniformly in the range 0$<${\tt R}$<$1, i.e. excluding the endpoints. \iteme{IDUMMY :} dummy input argument; normally 0. \end{entry} \drawbox{CALL RLUGET(LFN,MOVE)}\label{p:RLUGET} \begin{entry} \itemc{Purpose:} to dump the current state of the random number generator on a separate file, using internal representation for real and integer numbers. To be precise, the full contents of the {\tt LUDATR} common block are written on the file, with the exception of {\tt MRLU(6)}. \iteme{LFN :} (logical file number) the file number to which the state is dumped. You must associate this number with a true file (with a machine-dependent name), and see to it that this file is open for write. \iteme{MOVE :} choice of adding a new record to the file or overwriting old record(s). Normally only options 0 or $-$1 should be used. \begin{subentry} \iteme{= 0 (\mrm{or} > 0) :} add a new record to the end of the file. \iteme{= -1 :} overwrite the last record with a new one (i.e. do one {\tt BACKSPACE} before the new write). \iteme{= $-n$ :} back up $n$ records before writing the new record. The records following after the new one are lost, i.e. the last $n$ old records are lost and one new added. \end{subentry} \end{entry} \drawbox{CALL RLUSET(LFN,MOVE)}\label{p:RLUSET} \begin{entry} \itemc{Purpose:} to read in a state for the random number generator, from which the subsequent generation can proceed. The state must previously have been saved by a {\tt RLUGET} call. Again the full contents of the {\tt LUDATR} common block are read, with the exception of {\tt MRLU(6)}. \iteme{LFN :} (logical file number) the file number from which the state is read. You must associate this number with a true file previously written with a {\tt RLUGET} call, and see to it that this file is open for read. \iteme{MOVE :} positioning in file before a record is read. With zero value, records are read one after the other for each new call, while non-zero values may be used to navigate back and forth, and e.g. return to the same initial state several times. \begin{subentry} \iteme{= 0 :} read the next record. \iteme{= $+n$ :} skip ahead $n$ records before reading the record that sets the state of the random number generator. \iteme{= $-n$ :} back up $n$ records before reading the record that sets the state of the random number generator. \end{subentry} \end{entry} \drawbox{COMMON/LUDATR/MRLU(6),RRLU(100)}\label{p:LUDATR} \begin{entry} \itemc{Purpose:} to contain the state of the random number generator at any moment (for communication between {\tt RLU}, {\tt RLUGET} and {\tt RLUSET}), and also to provide the user with the possibility to initialize different random number sequences, and to know how many numbers have been generated. \iteme{MRLU(1) :}\label{p:MRLU} (D=19780503) the integer number that specifies which of the possible subsequences will be initialized in the next {\tt RLU} call for which {\tt MRLU(2)=0}. Allowed values are 0$\leq${\tt MRLU(1)}$\leq$900\,000\,000, the original Marsaglia (and CERN library) seed is 54217137. The {\tt MRLU(1)} value is not changed by any of the {\Je} routines. \iteme{MRLU(2) :} (D=0) initialization flag, put to 1 in the first {\tt RLU} call of run. A reinitialization of the random number generator can be made in mid-run by resetting {\tt MRLU(2)} to 0 by hand. In addition, any time the counter {\tt MRLU(3)} reaches 1000000000, it is reset to 0 and {\tt MRLU(2)} is increased by 1. \iteme{MRLU(3) :} (D=0) counter for the number of random numbers generated from the beginning of the run. To avoid overflow when very many numbers are generated, {\tt MRLU(2)} is used as described above. \iteme{MRLU(4), MRLU(5) :} {\tt I97} and {\tt J97} of the CERN library implementation; part of the state of the generator. \iteme{MRLU(6) :} (D=0) current position, i.e. how many records after beginning, in the file; used by {\tt RLUGET} and {\tt RLUSET}. \iteme{RRLU(1) - RRLU(97) :}\label{p:RRLU} the {\tt U} array of the CERN library implementation; part of the state of the generator. \iteme{RRLU(98) - RRLU(100) :} {\tt C}, {\tt CD} and {\tt CM} of the CERN library implementation; the first part of the state of the generator, the latter two constants calculated at initialization. \end{entry} \clearpage \section{The Event Record} The event record is the central repository for information about the particles produced in the current event: flavours, momenta, event history, and production vertices. It plays a very central r\^ole: without a proper understanding of what the record is and how information is stored, it is meaningless to try to use either {\Je} or {\Py}. The record is stored in the common block {\tt LUJETS}. Almost all the routines thatthe user calls can be viewed as performing some action on the record: fill a new event, let partons fragment or particles decay, boost it, list it, find clusters, etc. In this section we will first describe the KF flavour code, subsequently the {\tt LUJETS} common block, and then give a few comments about the r\^ole of the event record in the programs. To ease the interfacing of different event generators, a {\tt HEPEVT} standard common block structure for the event record has been agreed on. For historical reasons the standard common blocks are not directly used in {\Je}, but a conversion routine comes with the program, and is described at the end of this section. \subsection{Particle Codes} \label{ss:codes} The new particle code now adopted by the Particle Data Group \cite{PDG88,PDG92} is used consistently throughout the program, and is referred to as the KF particle code. This code you have to be thoroughly familiar with. It is described below. Note that a few inconsistencies between the KF and the PDG codes are known, which stem from differences of interpretation of the rules agreed on when developing the standard. These rules form the basis of the PDG tables and (independently) of the {\Je} tables. (Of course, my private opinion is that I follow the original agreement, and the PDG deviate from it.) Hopefully, this should have few practical consequences, since only rarely-produced particles are affected. Anyway, here is a list of the known discrepancies: \begin{Enumerate} \item The PDG has not allowed for the existence of an $\eta_{\b}$, which in {\Je} is included with code 551. This code is reserved for $\chi_{0 \b}$ by the PDG, a particle which appears as 10551 in {\Je}. (We agree to have $\eta_{\c}$ as 441, which illustrates the basic difference: I use the additional recurrence figure to refer to a whole multiplet, whether all particles of that multiplet have been found or not; the PDG, on the other hand, does not reserve space for particles which we know should be there but have not yet been discovered, which means that members of a multiplet need not go together.) \item The PDG has not allowed for the existence of an $\hrm_{1 \c}$, which in {\Je} is represented by 10443. Therefore $\chi_{1 \c}$ is the PDG code 10443 but {\Je} code 20443. Further $\psi'$ is either 20443 or 30443, and $\Upsilon' = \Upsilon(2S)$ either 20553 or 30553. (Comment as for point 1.) \item Different conventions for spin $1/2$ baryons with one heavy flavour (charm, bottom, top), one strange flavour, and one light ($\u$ or $\d$). Here two states exist, e.g. $\Xi_{\c}^+$ and $\Xi'^+_{\c}$, both with flavour content $\c\s\u$. By analogy with the $\Lambda^0$--$\Sigma$ pair, {\Je} uses the decreasing order of flavour content for the heavier state and inversed order of the two lighter flavours for the lighter state, while the PDG tables use the opposite convention. Thus in {\Je} $\Xi_{\c}^+$ is 4232 and $\Xi'^+_{\c}$ 4322, while in PDG it is the other way around. \end{Enumerate} There are no plans to change the {\Je} rules to agree with the PDG ones in either of the cases above. The KF code is not convenient for a direct storing of masses, decay data, or other particle properties, since the KF codes are so spread out. Instead a compressed code KC between 1 and 500 is used here, where the most frequently used particles have a separate code, but many heavy-flavour hadrons are lumped together in groups. Normally this code is only used at very specific places in the program, not visible to the user. If need be, the correspondence can always be obtained by using the function {\tt LUCOMP}, \mbox{\tt KC = LUCOMP(KF)}. It is therefore not intended that you should ever need to know any KC codes at all. It may be useful to know, however, that for codes smaller than 80, KF and KC agree. The particle names printed in the tables in this section correspond to the ones obtained with the routine {\tt LUNAME}, which is used extensively, e.g. in {\tt LULIST}. Greek characters are spelt out in full, with a capital first letter to correspond to a capital Greek letter. Generically the name of a particle is made up of the following pieces: \begin{Enumerate} \item The basic root name. This includes a * for most spin 1 ($L = 0$) mesons and spin $3/2$ baryons, and a $'$ for some spin $1/2$ baryons (where there are two states to be distinguished, cf. $\Lambda$--$\Sigma^0$). The rules for heavy baryon naming are in accordance with the 1986 Particle Data Group conventions \cite{PDG86}. For mesons with one unit of orbital angular momentum, K (D, B, \ldots) is used for quark-spin 0 and K* (D*, B*, \ldots) for quark-spin 1 mesons; the convention for `*' may here deviate slightly from the one used by the PDG. \item Any lower indices, separated from the root by a \_. For heavy hadrons, this is the additional heavy-flavour content not inherent in the root itself. For a diquark, it is the spin. \item The character $\sim$ (alternatively bar, see {\tt MSTU(15)}) for an antiparticle, wherever the distinction between particle and antiparticle is not inherent in the charge information. \item Charge information: $++$, $+$, $0$, $-$, or $--$. Charge is not given for quarks or diquarks. Some neutral particles which are customarily given without a 0 also here lack it, such as neutrinos, $\g$, $\gamma$, and flavour-diagonal mesons other than $\pi^0$ and $\rho^0$. Note that charge is included both for the proton and the neutron. While non-standard, it is helpful in avoiding misunderstandings when looking at an event listing. \end{Enumerate} Below follows a list of KF particle codes. The list is not complete; a more extensive one may be obtained with {\tt CALL LULIST(11)}. Particles are grouped together, and the basic rules are described for each group. Whenever a distinct antiparticle exists, it is given the same KF code with a minus sign (whereas KC codes are always positive). \begin{Enumerate} \item Quarks and leptons, Table \ref{t:codeone}. \\ This group contains the basic building blocks of matter, arranged according to family, with the lower member of weak isodoublets also having the smaller code (thus $\d$ precedes $\u$, contrary to the ordering in previous {\Je} versions). A fourth generation is included for future reference. The quark codes are used as building blocks for the diquark, meson and baryon codes below. \begin{table}[ptb] \captive{Quark and lepton codes. \protect\label{t:codeone} } \\ \vspace{1ex} \begin{center} \begin{tabular}{|c|c|c||c|c|c|@{\protect\rule{0mm}{\tablinsep}}} \hline KF & Name & Printed & KF & Name & Printed \\ \hline 1 & $\d$ & {\tt d} & 11 & $\e^-$ & {\tt e-} \\ 2 & $\u$ & {\tt u} & 12 & $\nu_{\e}$ & {\tt nu\_e} \\ 3 & $\s$ & {\tt s} & 13 & $\mu^-$ & {\tt mu-} \\ 4 & $\c$ & {\tt c} & 14 & $\nu_{\mu}$ & {\tt nu\_mu} \\ 5 & $\b$ & {\tt b} & 15 & $\tau^-$ & {\tt tau-} \\ 6 & $\t$ & {\tt t} & 16 & $\nu_{\tau}$ & {\tt nu\_tau} \\ 7 & $\lrm$ & {\tt l} & 17 & $\chi^-$ & {\tt chi-} \\ 8 & $\hrm$ & {\tt h} & 18 & $\nu_{\chi}$ & {\tt nu\_chi} \\ 9 & & & 19 & & \\ 10 & & & 20 & & \\ \hline \end{tabular} \end{center} \end{table} \item Gauge bosons and other fundamental bosons, Table \ref{t:codetwo}. \\ This group includes all the gauge and Higgs bosons of the standard model, as well as some of the bosons appearing in various extensions of it. The latter are not covered by the standard PDG codes. They correspond to one extra {\bf U(1)} group and one extra {\bf SU(2)} one, a further Higgs doublet, a (scalar, colour octet) techni-$\eta$, a (scalar) leptoquark $\L_{\Q}$, and a horizontal gauge boson $\R$ (coupling between families). Additionally, we here include the pomeron $\pomeron$ and reggeon $\reggeon$ `particles', which are important e.g. in the description of diffractive scattering, but have no obvious position anywhere in the classification scheme. \begin{table}[ptb] \captive{Gauge boson and other fundamental boson codes. \protect\label{t:codetwo} } \\ \vspace{1ex} \begin{center} \begin{tabular}{|c|c|c||c|c|c|@{\protect\rule{0mm}{\tablinsep}}} \hline KF & Name & Printed & KF & Name & Printed \\ \hline 21 & $\g$ & {\tt g} & 31 & & \\ 22 & $\gamma$ & {\tt gamma} & 32 & $\Z'^0$ & {\tt Z'0} \\ 23 & $\Z^0$ & {\tt Z0} & 33 & $\Z''^0$ & {\tt Z"0} \\ 24 & $\W^+$ & {\tt W+} & 34 & $\W'^+$ & {\tt W'+} \\ 25 & $\H^0$ & {\tt H0} & 35 & $\H'^0$ & {\tt H'0} \\ 26 & & & 36 & $\A^0$ & {\tt A0} \\ 27 & & & 37 & $\H^+$ & {\tt H+} \\ 28 & $\reggeon$ & {\tt reggeon} & 38 & $\eta_{\mrm{techni}}$ & {\tt eta\_tech0} \\ 29 & $\pomeron$ & {\tt pomeron} & 39 & $\L_{\Q}$ & {\tt LQ} \\ 30 & & & 40 & $\R^0$ & {\tt R0} \\ \hline \end{tabular} \end{center} \end{table} \item Free space. \\ The positions 41--80 are currently unused. In the future, they might come to be used, e.g. for supersymmetric partners of the particles above, or for some other kind of new physics. At the moment, they are at your disposal. \item Various special codes, Table \ref{t:codefour}. \\ In a Monte Carlo, it is always necessary to have codes that do not correspond to any specific particle, but are used to lump together groups of similar particles for decay treatment, or to specify generic decay products. These codes, which again are non-standard, are found between numbers 81 and 100. Several are not found in the event record, and therefore properly belong only to the KC group of codes. \begin{table}[ptb] \captive{Various special codes. \protect\label{t:codefour} } \\ \vspace{1ex} \begin{center} \begin{tabular}{|c|c|c|@{\protect\rule{0mm}{\tablinsep}}} \hline KF & Printed & Meaning \\ \hline 81 & {\tt specflav} & Spectator flavour; used in decay-product listings \\ 82 & {\tt rndmflav} & A random $\u$, $\d$, or $\s$ flavour; possible decay product \\ 83 & {\tt phasespa} & Simple isotropic phase-space decay \\ 84 & {\tt c-hadron} & Information on decay of generic charm hadron \\ 85 & {\tt b-hadron} & Information on decay of generic bottom hadron \\ 86 & {\tt t-hadron} & Information on decay of generic top hadron \\ 87 & {\tt l-hadron} & Information on decay of generic low hadron \\ 88 & {\tt h-hadron} & Information on decay of generic high hadron \\ 89 & {\tt Wvirt} & Off-mass-shell $\W$ in weak decays of $\t$, $\lrm$, $\hrm$ or $\chi$ \\ 90 & {\tt diquark} & Generic code for diquark colour information \\ 91 & {\tt cluster} & Parton system in cluster fragmentation \\ 92 & {\tt string} & Parton system in string fragmentation \\ 93 & {\tt indep.} & Parton system in independent fragmentation \\ 94 & {\tt CMshower} & Four-momentum of time-like showering system \\ 95 & {\tt SPHEaxis} & Event axis found with {\tt LUSPHE} \\ 96 & {\tt THRUaxis} & Event axis found with {\tt LUTHRU} \\ 97 & {\tt CLUSjet} & Jet (cluster) found with {\tt LUCLUS} \\ 98 & {\tt CELLjet} & Jet (cluster) found with {\tt LUCELL} \\ 99 & {\tt table} & Tabular output from {\tt LUTABU} \\ 100 & & \\ \hline \end{tabular} \end{center} \end{table} \item Diquark codes, Table \ref{t:codefive}. \\ A diquark made up of a quark with code $i$ and another with code $j$, where $i \geq j$, and with total spin $s$, is given the code \begin{equation} \mrm{KF} = 1000 i + 100 j + 2s + 1 ~, \end{equation} i.e. the tens position is left empty (cf. the baryon code below). Some of the most frequently used codes are listed in the table. All the lowest-lying spin 0 and 1 diquarks are included in the program. The corresponding KC code is 90, and it is mainly used to store colour charge. \begin{table}[ptb] \captive{Diquark codes. \protect\label{t:codefive} } \\ \vspace{1ex} \begin{center} \begin{tabular}{|c|c|c||c|c|c|@{\protect\rule{0mm}{\tablinsep}}} \hline KF & Name & Printed & KF & Name & Printed \\ \hline & & & 1103 & $\d\d_1$ & {\tt dd\_1} \\ 2101 & $\u\d_0$ & {\tt ud\_0} & 2103 & $\u\d_1$ & {\tt ud\_1} \\ & & & 2203 & $\u\u_1$ & {\tt uu\_1} \\ 3101 & $\s\d_0$ & {\tt sd\_0} & 3103 & $\s\d_1$ & {\tt sd\_1} \\ 3201 & $\s\u_0$ & {\tt su\_0} & 3203 & $\s\u_1$ & {\tt su\_1} \\ & & & 3303 & $\s\s_1$ & {\tt ss\_1} \\ \hline \end{tabular} \end{center} \end{table} \item Meson codes, Tables \ref{t:codesixa} and \ref{t:codesixb}. \\ A meson made up of a quark with code $i$ and an antiquark with code $-j$, $j \neq i$, and with total spin $s$, is given the code \begin{equation} \mrm{KF} = \left\{ 100 \max(i,j) + 10 \min(i,j) + 2s + 1 \right\} \, \mrm{sign}(i-j) \, (-1)^{\max(i,j)} ~. \end{equation} Note the presence of an extra $-$ sign if the heaviest quark is a down-type one. This is in accordance with the particle--antiparticle distinction adopted in the 1986 Review of Particle Properties \cite{PDG86}. It means for example that a $\B$ meson contains a $\bbar$ antiquark rather than a $\b$ quark. The flavour-diagonal states are arranged in order of ascending mass. The standard rule of having the last digit of the form $2s+1$ is broken for the $\K_{\mrm{S}}^0$--$\K_{\mrm{L}}^0$ system, where it is 0, and this convention should carry over to mixed states in the $\B$ meson system. For higher multiplets with the same spin, $\pm$10000, $\pm$20000, etc., are added to provide the extra distinction needed. Some of the most frequently used codes are given below. The full lowest-lying pseudoscalar and vector multiplets are included in the program, Table \ref{t:codesixa}. \begin{table}[ptb] \captive{Meson codes, part 1. \protect\label{t:codesixa} } \\ \vspace{1ex} \begin{center} \begin{tabular}{|c|c|c||c|c|c|@{\protect\rule{0mm}{\tablinsep}}} \hline KF & Name & Printed & KF & Name & Printed \\ \hline 211 & $\pi^+$ & {\tt pi+} & 213 & $\rho^+$ & {\tt rho+} \\ 311 & $\K^0$ & {\tt K0} & 313 & $\K^{*0}$ & {\tt K*0} \\ 321 & $\K^+$ & {\tt K+} & 323 & $\K^{*+}$ & {\tt K*+} \\ 411 & $\D^+$ & {\tt D+} & 413 & $\D^{*+}$ & {\tt D*+} \\ 421 & $\D^0$ & {\tt D0} & 423 & $\D^{*0}$ & {\tt D*0} \\ 431 & $\D_s^+$ & {\tt D\_s+} & 433 & $\D_{\s}^{*+}$ & {\tt D*\_s+} \\ 511 & $\B^0$ & {\tt B0} & 513 & $\B^{*0}$ & {\tt B*0} \\ 521 & $\B^+$ & {\tt B+} & 523 & $\B^{*+}$ & {\tt B*+} \\ 531 & $\B_s^0$ & {\tt B\_s0} & 533 & $\B_{\s}^{*0}$ & {\tt B*\_s0} \\ 541 & $\B_c^+$ & {\tt B\_c+} & 543 & $\B_{\c}^{*+}$ & {\tt B*\_c+} \\ 111 & $\pi^0$ & {\tt pi0} & 113 & $\rho^0$ & {\tt rho0} \\ 221 & $\eta$ & {\tt eta} & 223 & $\omega$ & {\tt omega} \\ 331 & $\eta'$ & {\tt eta'} & 333 & $\phi$ & {\tt phi} \\ 441 & $\eta_{\c}$ & {\tt eta\_c} & 443 & $\Jpsi$ & {\tt J/psi} \\ 551 & $\eta_{\b}$ & {\tt eta\_b} & 553 & $\Upsilon$ & {\tt Upsilon} \\ 661 & $\eta_{\t}$ & {\tt eta\_t} & 663 & $\Theta$ & {\tt Theta} \\ 130 & $\K_{\mrm{L}}^0$ & {\tt K\_L0} & & & \\ 310 & $\K_{\mrm{S}}^0$ & {\tt K\_S0} & & & \\ \hline \end{tabular} \end{center} \end{table} Also the lowest-lying orbital angular momentum $L = 1$ mesons are included, Table \ref{t:codesixb}: one pseudovector multiplet obtained for total quark-spin 0 ($L = 1, S = 0 \Rightarrow J = 1$) and one scalar, one pseudovector and one tensor multiplet obtained for total quark-spin 1 ($L = 1, S = 1 \Rightarrow J = 0, 1$ or 2), where $J$ is what is conventionally called the spin $s$ of the meson. Any mixing between the two pseudovector multiplets is not taken into account. Please note that some members of these multiplets have still not been found, and are included here only based on guesswork. Even for known ones, the information on particles (mass, width, decay modes) is highly incomplete. Only two radial excitations are included, the $\psi' = \psi(2S)$ and $\Upsilon' = \Upsilon(2S)$. \begin{table}[ptb] \captive{Meson codes, part 2. \protect\label{t:codesixb} } \\ \vspace{1ex} \begin{center} \begin{tabular}{|c|c|c||c|c|c|@{\protect\rule{0mm}{\tablinsep}}} \hline KF & Name & Printed & KF & Name & Printed \\ \hline 10213 & $\b_1$ & {\tt b\_1+} & 10211 & $\a_0^+$ & {\tt a\_0+} \\ 10313 & $\K_1^0$ & {\tt K\_10} & 10311 & $\K_0^{*0}$ & {\tt K*\_00} \\ 10323 & $\K_1^+$ & {\tt K\_1+} & 10321 & $\K_0^{*+}$ & {\tt K*\_0+} \\ 10413 & $\D_1^+$ & {\tt D\_1+} & 10411 & $\D_0^{*+}$ & {\tt D*\_0+} \\ 10423 & $\D_1^0$ & {\tt D\_10} & 10421 & $\D_0^{*0}$ & {\tt D*\_00} \\ 10433 & $\D_{1 \s}^+$ & {\tt D\_1s+} & 10431 & $\D_{0 \s}^{*+}$ & {\tt D*\_0s+} \\ 10113 & $\b_1^0$ & {\tt b\_10} & 10111 & $\a_0^0$ & {\tt a\_00} \\ 10223 & $\hrm_1^0$ & {\tt h\_10} & 10221 & $\f_0^0$ & {\tt f\_00} \\ 10333 & $\hrm'^0_1$ & {\tt h'\_10} & 10331 & $\f'^0_0$ & {\tt f'\_00} \\ 10443 & $\hrm_{1 \c}^0$ & {\tt h\_1c0} & 10441 & $\chi_{0 \c}^0$ & {\tt chi\_0c0} \\ \hline 20213 & $\a_1^+$ & {\tt a\_1+} & 215 & $\a_2^+$ & {\tt a\_2+} \\ 20313 & $\K_1^{*0}$ & {\tt K*\_10} & 315 & $\K_2^{*0}$ & {\tt K*\_20} \\ 20323 & $\K_1^{*+}$ & {\tt K*\_1+} & 325 & $\K_2^{*+}$ & {\tt K*\_2+} \\ 20413 & $\D_1^{*+}$ & {\tt D*\_1+} & 415 & $\D_2^{*+}$ & {\tt D*\_2+} \\ 20423 & $\D_1^{*0}$ & {\tt D*\_10} & 425 & $\D_2^{*0}$ & {\tt D*\_20} \\ 20433 & $\D_{1 \s}^{*+}$ & {\tt D*\_1s+} & 435 & $\D_{2 \s}^{*+}$ & {\tt D*\_2s+} \\ 20113 & $\a_1^0$ & {\tt a\_10} & 115 & $\a_2^0$ & {\tt a\_20} \\ 20223 & $\f_1^0$ & {\tt f\_10} & 225 & $\f_2^0$ & {\tt f\_20} \\ 20333 & $\f'^0_1$ & {\tt f'\_10} & 335 & $\f'^0_2$ & {\tt f'\_20} \\ 20443 & $\chi_{1 \c}^0$ & {\tt chi\_1c0} & 445 & $\chi_{2 \c}^0$ & {\tt chi\_2c0} \\ \hline 30443 & $\psi'$ & {\tt psi'} & & & \\ 30553 & $\Upsilon'$ & {\tt Upsilon'} & & & \\ \hline \end{tabular} \end{center} \end{table} The corresponding meson KC codes, used for organizing mass and decay data, range between 101 and 240. \item Baryon codes, Table \ref{t:codeseven}. \\ A baryon made up of quarks $i$, $j$ and $k$, with $i \geq j \geq k$, and total spin $s$, is given the code \begin{equation} \mrm{KF} = 1000 i + 100 j + 10 k + 2s + 1 ~. \end{equation} An exception is provided by spin $1/2$ baryons made up of three different types of quarks, where the two lightest quarks form a spin-0 diquark ($\Lambda$-like baryons). Here the order of the $j$ and $k$ quarks is reversed, so as to provide a simple means of distinction to baryons with the lightest quarks in a spin-1 diquark ($\Sigma$-like baryons). For hadrons with heavy flavours, the root names are Lambda or Sigma for hadrons with two $\u$ or $\d$ quarks, Xi for those with one, and Omega for those without $\u$ or $\d$ quarks. Some of the most frequently used codes are given in Table \ref{t:codeseven}. The full lowest-lying spin $1/2$ and $3/2$ multiplets are included in the program. The corresponding KC codes, used for organizing mass and decay data, range between 301 and 400, with some slots still free. \begin{table}[ptb] \captive{Baryon codes. \protect\label{t:codeseven} } \\ \vspace{1ex} \begin{center} \begin{tabular}{|c|c|c||c|c|c|@{\protect\rule{0mm}{\tablinsep}}} \hline KF & Name & Printed & KF & Name & Printed \\ \hline & & & 1114 & $\Delta^-$ & {\tt Delta-} \\ 2112 & $\n$ & {\tt n0} & 2114 & $\Delta^0$ & {\tt Delta0} \\ 2212 & $\p$ & {\tt p+} & 2214 & $\Delta^+$ & {\tt Delta+} \\ & & & 2224 & $\Delta^{++}$ & {\tt Delta++} \\ 3112 & $\Sigma^-$ & {\tt Sigma-} & 3114 & $\Sigma^{*-}$ & {\tt Sigma*-} \\ 3122 & $\Lambda^0$ & {\tt Lambda0} & & & \\ 3212 & $\Sigma^0$ & {\tt Sigma0} & 3214 & $\Sigma^{*0}$ & {\tt Sigma*0} \\ 3222 & $\Sigma^+$ & {\tt Sigma+} & 3224 & $\Sigma^{*+}$ & {\tt Sigma*+} \\ 3312 & $\Xi^-$ & {\tt Xi-} & 3314 & $\Xi^{*-}$ & {\tt Xi*-} \\ 3322 & $\Xi^0$ & {\tt Xi0} & 3324 & $\Xi^{*0}$ & {\tt Xi*0} \\ & & & 3334 & $\Omega^-$ & {\tt Omega-} \\ 4112 & $\Sigma_{\c}^0$ & {\tt Sigma\_c0} & 4114 & $\Sigma_{\c}^{*0}$ & {\tt Sigma*\_c0} \\ 4122 & $\Lambda_{\c}^+$ & {\tt Lambda\_c+} & & & \\ 4212 & $\Sigma_{\c}^+$ & {\tt Sigma\_c+} & 4214 & $\Sigma_{\c}^{*+}$ & {\tt Sigma*\_c+} \\ 4222 & $\Sigma_{\c}^{++}$ & {\tt Sigma\_c++} & 4224 & $\Sigma_{\c}^{*++}$ & {\tt Sigma*\_c++} \\ 4132 & $\Xi_{\c}^0$ & {\tt Xi\_c0} & & & \\ 4312 & $\Xi'^0_{\c}$ & {\tt Xi'\_c0} & 4314 & $\Xi_{\c}^{*0}$ & {\tt Xi*\_c0} \\ 4232 & $\Xi_{\c}^+$ & {\tt Xi\_c+} & & & \\ 4322 & $\Xi'^+_{\c}$ & {\tt Xi'\_c+} & 4324 & $\Xi_{\c}^{*+}$ & {\tt Xi*\_c+} \\ 4332 & $\Omega_{\c}^0$ & {\tt Omega\_c0} & 4334 & $\Omega_{\c}^{*0}$ & {\tt Omega*\_c0} \\ 5112 & $\Sigma_{\b}^-$ & {\tt Sigma\_b-} & 5114 & $\Sigma_{\b}^{*-}$ & {\tt Sigma*\_b-} \\ 5122 & $\Lambda_{\b}^0$ & {\tt Lambda\_b0} & & & \\ 5212 & $\Sigma_{\b}^0$ &{\tt Sigma\_b0} & 5214 & $\Sigma_{\b}^{*0}$ & {\tt Sigma*\_b0} \\ 5222 & $\Sigma_{\b}^+$ & {\tt Sigma\_b+} & 5224 & $\Sigma_{\b}^{*+}$ & {\tt Sigma*\_b+} \\ \hline \end{tabular} \end{center} \end{table} \item Diffractive states, Table \ref{t:codeeight}. \\ These codes are not standard ones: they have been defined by analogy to be used for denoting diffractive states in {\Py}, as part of the event history. The first two or three digits give flavour content, while the last one is 0, to denote the somewhat unusual character of the code. Only a few codes have been introduced; depending on circumstances these also have to double up for other diffractive states. \begin{table}[ptb] \captive{Diffractive state codes. \protect\label{t:codeeight} } \\ \vspace{1ex} \begin{center} \begin{tabular}{|c|c|c|@{\protect\rule{0mm}{\tablinsep}}} \hline KF & Printed & Meaning \\ \hline 110 & {\tt rho\_diff0} & Diffractive $\pi^0 / \rho^0 / \gamma$ state \\ 210 & {\tt pi\_diffr+} & Diffractive $\pi^+$ state \\ 220 & {\tt omega\_di0} & Diffractive $\omega$ state \\ 330 & {\tt phi\_diff0} & Diffractive $\phi$ state \\ 440 & {\tt J/psi\_di0} & Diffractive $\Jpsi$ state \\ 2110 & {\tt n\_diffr} & Diffractive $\n$ state \\ 2210 & {\tt p\_diffr+} & Diffractive $\p$ state \\ \hline \end{tabular} \end{center} \end{table} \item Free compressed codes. The positions 401--500 of mass and decay arrays are left open. Here a user may map any new kind of particle from the ordinary KF codes, which probably are above 10000, into a more manageable KC range for mass and decay data information. The mapping must be implemented in the {\tt LUCOMP} function. \end{Enumerate} \subsection{The Event Record} \label{ss:evrec} Each new event generated is in its entirety stored in the common block {\tt LUJETS}, which thus forms the event record. Here each jet or particle that appears at some stage of the fragmentation or decay chain will occupy one line in the matrices. The different components of this line will tell which jet/particle it is, from where it originates, its present status (fragmented/decayed or not), its momentum, energy and mass, and the space--time position of its production vertex. Note that {\tt K(I,3)}--{\tt K(I,5)} and the {\tt P} and {\tt V} vectors may take special meaning for some specific applications (e.g. sphericity or cluster analysis), as described in those connections. The event history information stored in {\tt K(I,3)}--{\tt K(I,5)} should not be taken too literally. In the particle decay chains, the meaning of a mother is well-defined, but the fragmentation description is more complicated. The primary hadrons produced in string fragmentation come from the string as a whole, rather than from an individual parton. Even when the string is not included in the history (see {\tt MSTU(16)}), the pointer from hadron to parton is deceptive. For instance, in a $\q\g\qbar$ event, those hadrons are pointing towards the $\q$ ($\qbar$) parton that were produced by fragmentation from that end of the string, according to the random procedure used in the fragmentation routine. No particles point to the $\g$. This assignment seldom agrees with the visual impression, and is not intended to. The common block {\tt LUJETS} has expanded with time, and can now house 4000 entries. This figure may seem ridiculously large, but actually the previous limit of 2000 was often reached in studies of high-$\pT$ processes at the LHC and SSC. This is because the event record contains not only the final particles, but also all intermediate partons and hadrons, which subsequenty showered, fragmented or decayed. Included are also a wealth of photons coming from $\pi^0$ decays; the simplest way of reducing the size of the event record is actually to switch off $\pi^0$ decays by {\tt MDCY(LUCOMP(111),1)=0}. Also note that some routines, such as {\tt LUCLUS} and {\tt LUCELL}, use memory after the event record proper as a working area. Still, to change the size of the common block, upwards or downwards, is easy: just do a global substitute in the common block and change the {\tt MSTU(4)} value to the new number. If more than 10000 lines are to be used, the packing of colour information should also be changed, see {\tt MSTU(5)}. \drawbox{COMMON/LUJETS/N,K(4000,5),P(4000,5),V(4000,5)}\label{p:LUJETS} \begin{entry} \itemc{Purpose:} to contain the event record, i.e. the complete list of all partons and particles in the current event. \iteme{N :}\label{p:N} number of lines in the {\tt K}, {\tt P} and {\tt V} matrices occupied by the current event. {\tt N} is continuously updated as the definition of the original configuration and the treatment of fragmentation and decay proceed. In the following, the individual parton/particle number, running between 1 and {\tt N}, is called {\tt I}. \iteme{K(I,1) :}\label{p:K} status code KS, which gives the current status of the parton/particle stored in the line. The ground rule is that codes 1--10 correspond to currently existing partons/particles, while larger codes contain partons/particles which no longer exist, or other kinds of event information. \begin{subentry} \iteme{= 0 :} empty line. \iteme{= 1 :} an undecayed particle or an unfragmented jet, the latter being either a single jet or the last one of a jet system. \iteme{= 2 :} an unfragmented jet, which is followed by more jets in the same colour-singlet jet system. \iteme{= 3 :} an unfragmented jet with special colour flow information stored in {\tt K(I,4)} and {\tt K(I,5)}, such that adjacent partons along the string need not follow each other in the event record. \iteme{= 4 :} a particle which could have decayed, but did not within the allowed volume around the original vertex. \iteme{= 5 :} a particle which is to be forced to decay in the next {\tt LUEXEC} call, in the vertex position given (this code is only set by user intervention). \iteme{= 11 :} a decayed particle or a fragmented jet, the latter being either a single jet or the last one of a jet system, cf. {\tt =1}. \iteme{= 12 :} a fragmented jet, which is followed by more jets in the same colour-singlet jet system, cf. {\tt =2}. Further, a $\B$ meson which decayed as a $\br{\B}$ one, or vice versa, because of $\B$--$\br{\B}$ mixing, is marked with this code rather than {\tt =11}. \iteme{= 13 :} a jet which has been removed when special colour flow information has been used to rearrange a jet system, cf. {\tt =3}. \iteme{= 14 :} a parton which has branched into further partons, with special colour-flow information provided, cf. {\tt =3}. \iteme{= 15 :} a particle which has been forced to decay (by user intervention), cf. {\tt =5}. \iteme{= 21 :} documentation lines used to give a compressed story of the event at the beginning of the event record. \iteme{= 31 :} lines with information on sphericity, thrust or cluster search. \iteme{= 32 :} tabular output, as generated by {\tt LUTABU}. \iteme{= 41 :} junction (currently not fully implemented). \iteme{< 0 :} these codes are never used by the program, and are therefore usually not affected by operations on the record, such as {\tt LUROBO}, {\tt LULIST} and event-analysis routines (the exception is some {\tt LUEDIT} calls, where lines are moved but not deleted). Such codes may therefore be useful in some connections. \end{subentry} \iteme{K(I,2) :} parton/particle KF code, as described in section \ref{ss:codes}. \iteme{K(I,3) :} line number of parent particle or jet, where known, otherwise 0. Note that the assignment of a particle to a given jet in a jet system is unphysical, and what is given there is only related to the way the event was generated. \iteme{K(I,4) :} normally the line number of the first daughter; it is 0 for an undecayed particle or unfragmented jet. For {\tt K(I,1) = 3, 13} or {\tt 14}, instead, it contains special colour-flow information (for internal use only) of the form \\ {\tt K(I,4)} = 200000000*MCFR + 100000000*MCTO + 10000*ICFR + ICTO, \\ where ICFR and ICTO give the line numbers of the partons from which the colour comes and to where it goes, respectively; MCFR and MCTO originally are 0 and are set to 1 when the corresponding colour connection has been traced in the {\tt LUPREP} rearrangement procedure. (The packing may be changed with {\tt MSTU(5)}.) The `from' colour position may indicate a parton which branched to produce the current parton, or a parton created together with the current parton but with matched anticolour, while the `to' normally indicates a parton that the current parton branches into. Thus, for setting up an initial colour configuration, it is normally only the `from' part that is used, while the `to' part is added by the program in a subsequent call to parton-shower evolution (for final-state radiation; it is the other way around for initial-state radiation). {\bf Note:} normally most users never have to worry about the exact rules for colour-flow storage, since this is used mainly for internal purposes. However, when it is necessary to define this flow, it is recommended to use the {\tt LUJOIN} routine, since it is likely that this would reduce the chances of making a mistake. \iteme{K(I,5) :} normally the line number of the last daughter; it is 0 for an undecayed particle or unfragmented jet. For {\tt K(I,1) = 3, 13} or {\tt 14}, instead, it contains special colour-flow information (for internal use only) of the form \\ {\tt K(I,5)} = 200000000*MCFR + 100000000*MCTO + 10000*ICFR + ICTO, \\ where ICFR and ICTO give the line numbers of the partons from which the anticolour comes and to where it goes, respectively; MCFR and MCTO originally are 0 and are set to 1 when the corresponding colour connection has been traced in the {\tt LUPREP} rearrangement procedure. For further discussion, see {\tt K(I,4)}. \iteme{P(I,1) :}\label{p:P} $p_x$, momentum in the $x$ direction, in GeV/$c$. \iteme{P(I,2) :} $p_y$, momentum in the $y$ direction, in GeV/$c$. \iteme{P(I,3) :} $p_z$, momentum in the $z$ direction, in GeV/$c$. \iteme{P(I,4) :} $E$, energy, in GeV. \iteme{P(I,5) :} $m$, mass, in GeV/$c^2$. In parton showers, with space-like virtualities, i.e. where $Q^2 = - m^2 > 0$, one puts {\tt P(I,5)}$ = -Q$. \iteme{V(I,1) :}\label{p:V} $x$ position of production vertex, in mm. \iteme{V(I,2) :} $y$ position of production vertex, in mm. \iteme{V(I,3) :} $z$ position of production vertex, in mm. \iteme{V(I,4) :} time of production, in mm/$c$ ($\approx 3.33 \times 10^{-12}$ s). \iteme{V(I,5) :} proper lifetime of particle, in mm/$c$ ($\approx 3.33 \times 10^{-12}$ s). If the particle is not expected to decay, {\tt V(I,5)=0}. A line with {\tt K(I,1)=4}, i.e. a particle that could have decayed, but did not within the allowed region, has the proper non-zero {\tt V(I,5)}. In the absence of electric or magnetic fields, or other disturbances, the decay vertex {\tt VP} of an unstable particle may be calculated as \\ {\tt VP(j) = V(I,j) + V(I,5)*P(I,j)/P(I,5)}, {\tt j} = 1--4. \end{entry} \subsection{How The Event Record Works} The event record is the main repository for information about an event. In the generation chain, it is used as a `scoreboard' for what has already been done and what remains to be done. This information can be studied by you, to access information not only about the final state, but also about what came before. \subsubsection{A simple example} The example of section \ref{ss:JETstarted} may help to clarify what is going on. When {\tt LU2ENT} is called to generate a $\q\qbar$ pair, the quarks are stored in lines 1 and 2 of the event record, respectively. Colour information is set to show that they belong together as a colour singlet. The counter {\tt N} is also updated to the value of 2. At no stage is the previously generated event removed. Lines 1 and 2 are overwritten, but lines 3 onwards still contain whatever may have been there before. This does not matter, since {\tt N} indicates where the `real' record ends. As {\tt LUEXEC} is called, explicitly by you or indirectly by {\tt LU2ENT}, the first entry is considered and found to be the first jet of a system. Therefore the second entry is also found, and these two together form a jet system, which may be allowed to fragment. The `string' that fragments is put in line 3 and the fragmentation products in lines 4 through 10 (in this particular case). At the same time, the $\q$ and $\qbar$ in the first two lines are marked as having fragmented, and the same for the string. At this stage, {\tt N} is 10. Internally there is another counter with the value 2, which indicates how far down in the record the event has been studied. This second counter is gradually increased by one. If the entry in the corresponding line can fragment or decay, then fragmentation or decay is perfomed. The fragmentation/decay products are added at the end of the event record, and {\tt N} is updated accordingly. The entry is then also marked as having been treated. For instance, when line 3 is considered, the `string' entry of this line is seen to have been fragmented, and no action is taken. Line 4, a $\rho^+$, is allowed to decay to $\pi^+ \pi^0$; the decay products are stored in lines 11 and 12, and line 4 is marked as having decayed. Next, entry 5 is allowed to decay. The entry in line 6, $\pi^+$, is a stable particle (by default) and is therefore passed by without any action being taken. In the beginning of the process, entries are usually unstable, and {\tt N} grows faster than the second counter of treated entries. Later on, an increasing fraction of the entries are stable end products, and the r\^oles are now reversed, with the second counter growing faster. When the two coincide, the end of the record has been reached, and the process can be stopped. All unstable objects have now been allowed to fragment or decay. They are still present in the record, so as to simplify the tracing of the history. Notice that {\tt LUEXEC} could well be called a second time. The second counter would then start all over from the beginning, but slide through until the end without causing any action, since all objects that can be treated already have been. Unless some of the relevant switches were changed meanwhile, that is. For instance, if $\pi^0$ decays were switched off the first time around but on the second, all the $\pi^0$'s found in the record would be allowed to decay in the second call. A particle once decayed is not `undecayed', however, so if the $\pi^0$ is put back stable and {\tt LUEXEC} is called a third time, nothing will happen. \subsubsection{Applications to PYTHIA} \label{sss:PYrecord} In a full-blown event generated with {\Py}, the usage of {\tt LUJETS} is more complicated, although the general principles survive. {\tt LUJETS} is used extensively both by the {\Py} and the {\Je} routines; indeed it provides the bridge that allows the general utility routines in {\Je} to be used also for {\Py} events. The {\Py} event listing begins (optionally) with a few lines of event summary, specific to the hard process simulated and thus not described in the overview above. These specific parts are covered in the following. In most instances, only the partons and particles actually produced are of interest. For {\tt MSTP(125)=0}, the event record starts off with the parton configuration existing after hard interaction, initial- and final-state radiation, multiple interactions and beam remnants have been considered. The partons are arranged in colour singlet clusters, ordered as required for string fragmentation. Also photons and leptons produced as part of the hard interaction (e.g. from $\q\qbar \to \g \gamma$ or $\u\ubar \to \Z^0 \to \ee$) appear in this part of the event record. These original entries appear with pointer {\tt K(I,3)=0}, whereas the products of the subsequent fragmentation and decay have {\tt K(I,3)} numbers pointing back to the line of the parent. The standard documentation, obtained with {\tt MSTP(125)=1}, includes a few lines at the beginning of the event record, which contain a brief summary of the process that has taken place. The number of lines used depends on the nature of the hard process and is stored in {\tt MSTI(4)} for the current event. These lines all have {\tt K(I,1)=21}. For all processes, lines 1 and 2 give the two incoming hadrons. When listed with {\tt LULIST}, these two lines will be separated from subsequent ones by a sequence of `{\tt ======}' signs, to improve readability. For diffractive and elastic events, the two outgoing states in lines 3 and 4 complete the list. Otherwise, lines 3 and 4 contain the two partons that initiate the two initial-state parton showers, and 5 and 6 the end products of these showers, i.e. the partons that enter the hard interaction. With initial-state radiation switched off, lines 3 and 5 and lines 4 and 6 coincide. For a simple $2 \to 2$ hard scattering, lines 7 and 8 give the two outgoing partons/particles from the hard interaction, before any final-state radiation. For $2 \to 2$ processes proceeding via an intermediate resonance such as $\gammaZ$, $\W^{\pm}$ or $\H^0$, the resonance is found in line 7 and the two outgoing partons/particles in 8 and 9. In some cases one of these may be a resonance in its own right, or both of them, so that further pairs of lines are added for subsequent decays. If the decay of a given resonance has been switched off, then no decay products are listed either in this initial summary or in the subsequent ordinary listing. Whenever partons are listed, they are assumed to be on the mass shell for simplicity. The fact that effective masses may be generated by initial- and final-state radiation is taken into account in the actual parton configuration that is allowed to fragment, however. A special case is provided by $\W^+ \W^-$ or $\Z^0 \Z^0$ fusion to an $\H^0$. Then the virtual $\W$'s or $\Z$'s are shown in lines 7 and 8, the $\H^0$ in line 9, and the two recoiling quarks (that emitted the bosons) in 10 and 11, followed by the Higgs decay products. Since the $\W$'s and $\Z$'s are space-like, what is actually listed as the mass for them is $-\sqrt{-m^2}$. The listing of the event documentation closes with another line made up of `{\tt ======}' signs. A few examples may help clarify the picture. For a single diffractive event $\p \pbar \to \p_{\mrm{diffr}} \pbar$, the event record will start with \\ \verb& I K(I,1) K(I,2) K(I,3) & comment \\ \verb& 1 21 2212 0 & incoming $\p$ \\ \verb& 2 21 -2212 0 & incoming $\pbar$ \\ \verb&========================= & not part of record; appears in listings \\ \verb& 3 21 27 1 & outgoing $\p_{\mrm{diffr}}$ \\ \verb& 4 21 -2212 2 & outgoing $\pbar$ \\ \verb&========================= & again not part of record The typical QCD $2 \to 2$ process would be \\ \verb& I K(I,1) K(I,2) K(I,3) & comment \\ \verb& 1 21 2212 0 & incoming $\p$ \\ \verb& 2 21 -2212 0 & incoming $\pbar$ \\ \verb&========================= & \\ \verb& 3 21 2 1 & $\u$ picked from incoming $\p$ \\ \verb& 4 21 -1 2 & $\dbar$ picked from incoming $\pbar$ \\ \verb& 5 21 21 3 & $\u$ evolved to $\g$ at hard scattering \\ \verb& 6 21 -1 4 & still $\dbar$ at hard scattering \\ \verb& 7 21 21 0 & outgoing $\g$ from hard scattering \\ \verb& 8 21 -1 0 & outgoing $\dbar$ from hard scattering \\ \verb&========================= & Note that, where well defined, the {\tt K(I,3)} code does contain information as to which side the different partons come from, e.g. above the gluon in line 5 points back to the $\u$ in line 3, which points back to the proton in line 1. In the example above, it would have been possible to associate the scattered g in line 7 with the incoming one in line 5, but this is not possible in the general case, consider e.g. $\g \g \to \g \g$. As a final example, $\W^+\W^-$ fusion to an $\H^0$ might look like \\ \verb& I K(I,1) K(I,2) K(I,3) & comment \\ \verb& 1 21 2212 0 & first incoming $\p$ \\ \verb& 2 21 2212 0 & second incoming $\p$ \\ \verb&========================= & \\ \verb& 3 21 2 1 & $\u$ picked from first $\p$ \\ \verb& 4 21 21 2 & $\g$ picked from second $\p$ \\ \verb& 5 21 2 3 & still $\u$ after initial-state radiation \\ \verb& 6 21 -4 4 & $\g$ evolved to $\cbar$ \\ \verb& 7 21 24 5 & space-like $\W^+$ emitted by $\u$ quark \\ \verb& 8 21 -24 6 & space-like $\W^-$ emitted by $\cbar$ quark \\ \verb& 9 21 25 0 & Higgs produced by $\W^+ \W^-$ fusion \\ \verb&10 21 1 5 & $\u$ turned into $\d$ by emission of $\W^+$ \\ \verb&11 21 -3 6 & $\cbar$ turned into $\sbar$ by emission of $\W^-$ \\ \verb&12 21 23 9 & first $\Z^0$ coming from decay of $\H^0$ \\ \verb&13 21 23 9 & second $\Z^0$ coming from decay of $\H^0$ \\ \verb&14 21 12 12 & $\nu_{\e}$ from first $\Z^0$ decay \\ \verb&15 21 -12 12 & $\br{\nu}_{\e}$ from first $\Z^0$ decay \\ \verb&16 21 5 13 & $\b$ quark from second $\Z^0$ decay \\ \verb&17 21 -5 13 & $\bbar$ antiquark from second $\Z^0$ decay \\ \verb&========================= & After these lines with the initial information, the event record looks the same as for {\tt MSTP(125)=0}, i.e. first comes the parton configuration to be fragmented and, after another separator line `{\tt ======}' in the output (but not the event record), the products of subsequent fragmentation and decay chains. The {\tt K(I,3)} pointers for the partons, as well as leptons and photons produced in the hard interaction, are now pointing towards the documentation lines above, however. In particular, beam remnants point to 1 or 2, depending on which side they belong to, and partons emitted in the initial-state parton showers point to 3 or 4. In the second example above, the partons produced by final-state radiation will be pointing back to 7 and 8; as usual, it should be remembered that a specific assignment to 7 or 8 need not be unique. For the third example, final-state radiation partons will come both from partons 10 and 11 and from partons 16 and 17, and additionally there will be a neutrino--antineutrino pair pointing to 14 and 15. The extra pairs of partons that are generated by multiple interactions do not point back to anything, i.e. they have {\tt K(I,3)=0}. There exists a third documentation option, {\tt MSTP(125)=2}. Here the history of initial- and final-state parton branchings may be traced, including all details on colour flow. This information has not been optimized for user-friendliness, and cannot be recommended for general usage. With this option, the initial documentation lines are the same. They are followed by blank lines, {\tt K(I,1)=0}, up to line 20 (can be changed in {\tt MSTP(126)}). From line 21 onwards each parton with {\tt K(I,1)=} 3, 13 or 14 appears with special colour-flow information in the {\tt K(I,4)} and {\tt K(I,5)} positions. For an ordinary $2 \to 2$ scattering, the two incoming partons at the hard scattering are stored in lines 21 and 22, and the two outgoing in 23 and 24. The colour flow between these partons has to be chosen according to the proper relative probabilities in cases when many alternatives are possible, see section \ref{sss:QCDjetclass}. If there is initial-state radiation, the two partons in lines 21 and 22 are copied down to lines 25 and 26, from which the initial-state showers are reconstructed backwards step by step. The branching history may be read by noting that, for a branching $a \to b c$, the {\tt K(I,3)} codes of $b$ and $c$ point towards the line number of $a$. Since the showers are reconstructed backwards, this actually means that parton $b$ would appear in the listing before parton $a$ and $c$, and hence have a pointer to a position below itself in the list. Associated time-like partons $c$ may initiate time-like showers, as may the partons of the hard scattering. Again a showering parton or pair of partons will be copied down towards the end of the list and allowed to undergo successive branchings $c \to d e$, with $d$ and $e$ pointing towards $c$. The mass of time-like partons is properly stored in {\tt P(I,5)}; for space-like partons $-\sqrt{-m^2}$ is stored instead. After this section, containing all the branchings, comes the final parton configuration, properly arranged in colour, followed by all subsequent fragmentation and decay products, as usual. \subsection{The HEPEVT Standard} \label{ss:HEPEVT} A set of common blocks was developed and agreed on within the framework of the 1989 LEP physics study, see \cite{Sjo89}. This standard defines an event record structure which should make the interfacing of different event generators much simpler. It would be a major work to rewrite {\PyJe} to agree with this standard event record structure. More importantly, the standard only covers quantities which can be defined unambiguously, i.e. which are independent of the particular program used. There are thus no provisions for the need for colour-flow information in models based on string fragmentation, etc., so the standard common blocks would anyway have to be supplemented with additional event information. For the moment, the adopted approach is therefore to retain the {\tt LUJETS} event record, but supply a routine {\tt LUHEPC} which can convert to or from the standard event record. Owing to a somewhat different content in the two records, some ambiguities do exist in the translation procedure. {\tt LUHEPC} has therefore to be used with some judgment. In this section, the new standard event structure is first presented, i.e. the most important points in \cite{Sjo89} are recapitulated. Thereafter the conversion routine is described, with particular attention to ambiguities and limitations. The standard event record is stored in two common blocks. The second of these is specifically intended for spin information. Since {\Je} never (explicitly) makes use of spin information, this latter common block is not addressed here. In order to make the components of the standard more distinguishable in user programs, the three characters {\tt HEP} (for High Energy Physics) have been chosen to be a part of all names. \drawboxthree{~PARAMETER (NMXHEP=2000)} {~COMMON/HEPEVT/NEVHEP,NHEP,ISTHEP(NMXHEP),IDHEP(NMXHEP),} {\&JMOHEP(2,NMXHEP),JDAHEP(2,NMXHEP),PHEP(5,NMXHEP),VHEP(4,NMXHEP)} \label{p:HEPEVT}\begin{entry} \itemc{Purpose:} to contain an event record in a Monte Carlo-independent format. \iteme{NMXHEP:} maximum numbers of entries (partons/particles) that can be stored in the common block. The default value of 2000 can be changed via the parameter construction. In the translation, it is checked that this value is not exceeded. \iteme{NEVHEP:} is normally the event number, but may have special meanings, according to the description below: \begin{subentry} \iteme{> 0 :} event number, sequentially increased by 1 for each call to the main event generation routine, starting with 1 for the first event generated. \iteme{= 0 :} for a program which does not keep track of event numbers, as {\Je}. \iteme{= -1 :} special initialization record; not used by {\Je}. \iteme{= -2 :} special final record; not used by {\Je}. \end{subentry} \iteme{NHEP:} the actual number of entries stored in the current event. These are found in the first {\tt NHEP} positions of the respective arrays below. Index {\tt IHEP}, 1$\leq${\tt IHEP}$\leq${\tt NHEP}, is used below to denote a given entry. \iteme{ISTHEP(IHEP):} status code for entry {\tt IHEP}, with the following meanings: \begin{subentry} \iteme{= 0 :} null entry. \iteme{= 1 :} an existing entry, which has not decayed or fragmented. This is the main class of entries, which represents the `final state' given by the generator. \iteme{= 2 :} an entry which has decayed or fragmented and is therefore not appearing in the final state, but is retained for event history information. \iteme{= 3 :} a documentation line, defined separately from the event history. This could include the two incoming reacting particles, etc. \iteme{= 4 - 10 :} undefined, but reserved for future standards. \iteme{= 11 - 200 :} at the disposal of each model builder for constructs specific to his program, but equivalent to a null line in the context of any other program. \iteme{= 201 - :} at the disposal of users, in particular for event tracking in the detector. \end{subentry} \iteme{IDHEP(IHEP) :} particle identity, according to the PDG standard. The four additional codes 91--94 have been introduced to make the event history more legible, see section \ref{ss:codes} and the {\tt MSTU(16)} description. \iteme{JMOHEP(1,IHEP) :} pointer to the position where the mother is stored. The value is 0 for initial entries. \iteme{JMOHEP(2,IHEP) :} pointer to position of second mother. Normally only one mother exists, in which case the value 0 is to be used. In {\Je}, entries with codes 91--94 are the only ones to have two mothers. The flavour contents of these objects, as well as details of momentum sharing, have to be found by looking at the mother partons, i.e. the two partons in positions {\tt JMOHEP(1,IHEP)} and {\tt JMOHEP(2,IHEP)} for a cluster or a shower system, and the range {\tt JMOHEP(1,IHEP)}--{\tt JMOHEP(2,IHEP)} for a string or an independent fragmentation parton system. \iteme{JDAHEP(1,IHEP) :} pointer to the position of the first daughter. If an entry has not decayed, this is 0. \iteme{JDAHEP(2,IHEP) :} pointer to the position of the last daughter. If an entry has not decayed, this is 0. It is assumed that daughters are stored sequentially, so that the whole range {\tt JDAHEP(1,IHEP)}--{\tt JDAHEP(2,IHEP)} contains daughters. This variable should be set also when only one daughter is present, as in $\K^0 \to \K_{\mrm{S}}^0$ decays, so that looping from the first daughter to the last one works transparently. Normally daughters are stored after mothers, but in backwards evolution of initial-state radiation the opposite may appear, i.e. that mothers are found below the daughters they branch into. Also, the two daughters then need not appear one after the other, but may be separated in the event record. \iteme{PHEP(1,IHEP) :} momentum in the $x$ direction, in GeV/$c$. \iteme{PHEP(2,IHEP) :} momentum in the $y$ direction, in GeV/$c$. \iteme{PHEP(3,IHEP) :} momentum in the $z$ direction, in GeV/$c$. \iteme{PHEP(4,IHEP) :} energy, in GeV. \iteme{PHEP(5,IHEP) :} mass, in GeV/$c^2$. For space-like partons, it is allowed to use a negative mass, according to {\tt PHEP(5,IHEP)}$ = -\sqrt{-m^2}$. \iteme{VHEP(1,IHEP) :} production vertex $x$ position, in mm. \iteme{VHEP(2,IHEP) :} production vertex $y$ position, in mm. \iteme{VHEP(3,IHEP) :} production vertex $z$ position, in mm. \iteme{VHEP(4,IHEP) :} production time, in mm/$c$ ($\approx 3.33 \times 10^{-12}$ s). \end{entry} \boxsep This completes the brief description of the standard. In {\Je}, the routine {\tt LUHEPC} is provided as an interface. \drawbox{CALL LUHEPC(MCONV)}\label{p:LUHEPC} \begin{entry} \itemc{Purpose:} to convert between the {\tt LUJETS} event record and the {\tt HEPEVT} event record. \iteme{MCONV :} direction of conversion. \begin{subentry} \iteme{= 1 :} translates the current {\tt LUJETS} record into the {\tt HEPEVT} one, while leaving the original {\tt LUJETS} one unaffected. \iteme{= 2 :} translates the current {\tt HEPEVT} record into the {\tt LUJETS} one, while leaving the original {\tt HEPEVT} one unaffected. \end{subentry} \end{entry} \boxsep The conversion of momenta is trivial: it is just a matter of exchanging the order of the indices. The vertex information is but little more complicated; the extra fifth component present in {\tt LUJETS} can be easily reconstructed from other information for particles which have decayed. (Some of the advanced features made possible by this component, such as the possibility to consider decays within expanding spatial volumes in subsequent {\tt LUEXEC} calls, cannot be used if the record is translated back and forth, however.) Also, the particle codes {\tt K(I,2)} and {\tt IDHEP(I)} are identical, since they are both based on the PDG codes. The remaining, non-trivial areas deal with the status codes and the event history. In moving from {\tt LUJETS} to {\tt HEPEVT}, information on colour flow is lost. On the other hand, the position of a second mother, if any, has to be found; this only affects lines with {\tt K(I,2)=} 91--94. Also, for lines with {\tt K(I,1)=} 13 or 14, the daughter pointers have to be found. By and large, however, the translation from {\tt LUJETS} to {\tt HEPEVT} should cause little problem, and there should never be any need for user intervention. (We assume that {\Je} is run with the default {\tt MSTU(16)=1}, otherwise some discrepancies with respect to the proposed standard event history description will be present.) In moving from {\tt HEPEVT} to {\tt LUJETS}, information on a second mother is lost. Any codes {\tt IDHEP(I)} not equal to 1, 2 or 3 are translated into {\tt K(I,1)=0}, and so all entries with {\tt K(I,1)}$\geq 30$ are effectively lost in a translation back and forth. All entries with {\tt IDHEP(I)=2} are translated into {\tt K(I,1)=11}, and so entries of type {\tt K(I,1) = 12, 13, 14} or {\tt 15} are never found. There is thus no colour-flow information available for partons which have fragmented. For partons with {\tt IDHEP(I)=1}, i.e. which have not fragmented, an attempt is made to subdivide the partonic system into colour singlets, as required for subsequent string fragmentation. To this end, it is assumed that partons are stored sequentially along strings. Normally, a string would then start at a $\q$ ($\qbar$) or $\qbar\qbar$ ($\q\q$) entry, cover a number of intermediate gluons, and end at a $\qbar$ ($\q$) or $\q\q$ ($\qbar\qbar$) entry. Particles could be interspersed in this list with no adverse effects, i.e. a $\u-\g-\gamma-\ubar$ sequence would be interpreted as a $\u-\g-\ubar$ string plus an additional photon. A closed gluon loop would be assumed to be made up of a sequential listing of the gluons, with the string continuing from the last gluon up back to the first one. Contrary to the previous, open string case, the appearance of any particle but a gluon would therefore signal the end of the gluon loop. For example, a $\g-\g-\g-\g$ sequence would be interpreted as one single four-gluon loop, while a $\g-\g-\gamma-\g-\g$ sequence would be seen as composed of two 2-gluon systems. If these interpretations, which are not unique, are not to your liking, it is up to you to correct them, e.g. by using {\tt LUJOIN} to tell exactly which partons should be joined, in which sequence, to give a string. Calls to {\tt LUJOIN} (or the equivalent) are also necessary if {\tt LUSHOW} is to be used to have some partons develop a shower. For practical applications, one should note that {\Je} $\ee$ events, which have been allowed to shower but not to fragment, do have partons arranged in the order assumed above, so that a translation to {\tt HEPEVT} and back does not destroy the possibility to perform fragmentation by a simple {\tt LUEXEC} call. Also the hard interactions in {\Py} fulfil this condition, while problems may appear in the multiple interaction scenario, where several closed $\g\g$ loops may appear directly following one another, and thus would be interpreted as a single multigluon loop after translation back and forth. \clearpage \section{Hard Processes in JETSET} \label{s:JETSETproc} {\Je} contains the simulation of two hard processes. The process of main interest is $\ee \to \gammaZ \to \q \qbar$. Higher-order QCD corrections can be obtained either with parton showers or with second-order matrix elements. The details of the parton-shower evolution are given in section \ref{s:showinfi}, while this section contains the matrix-element description, including a summary of the {\Je} algorithm for initial-state photon radiation. Also {\Py} can be used to simulate the process $\ee \to \gammaZ \to \q \qbar$, but without the options of using second-order matrix elements or polarized incoming beams. Some other differences between the two algorithms are described. The other hard process in {\Je} is $\Upsilon$ decay to $\g \g \g$ or $\gamma \g \g$, which is briefly commented on. The main sources of information for this chapter are refs. \cite{Sjo83,Sjo86,Sjo89}. \subsection{Annihilation Events in the Continuum} \label{ss:eematrix} The description of $\ee$ annihilation into hadronic events involves a number of components: the $s$ dependence of the total cross section and flavour composition, multijet matrix elements, angular orientation of events, initial-state photon bremsstrahlung and effects of initial-state electron polarization. Many of the published formulae have been derived for the case of massless outgoing quarks. For each of the components described in the following, we will begin by discussing the massless case, and then comment on what is done to accommodate massive quarks. \subsubsection{Electroweak cross sections} In the standard theory, fermions have the following couplings (illustrated here for the first generation): \begin{center} \begin{tabular}{lll@{\protect\rule{0mm}{\tablinsep}}} $e_{\nu} = 0$, & $v_{\nu} = 1$, & $a_{\nu} = 1$, \\ $e_{\e} = -1$, & $v_{\e} = -1 + 4\ssintw$, & $a_{\e} = -1$, \\ $e_{\u} = 2/3$, & $v_{\u} = 1 - 8\ssintw /3$, & $a_{\nu} = 1$, \\ $e_{\d} = -1/3$, & $v_{\d} = -1 + 4\ssintw /3$, & $a_{\d} = -1$, \\ \end{tabular} \end{center} with $e$ the electric charge, and $v$ and $a$ the vector and axial couplings to the $\Z^0$. The relative energy dependence of the weak neutral current to the electromagnetic one is given by \begin{equation} \chi(s) = \frac{1}{4\ssintw\scostw} \; \frac{s}{s - m_{\Z}^2 + i m_{\Z}\Gamma_{\Z}} ~, \label{ee:chis} \end{equation} where $s = E_{\mrm{cm}}^2$. In {\Je} the electroweak mixing parameter $\ssintw$ and the $\Z^0$ mass $m_{\Z}$ and width $\Gamma_{\Z}$ are considered as constants to be given by you (while {\Py} itself calculates an $s$-dependent width). Although the incoming $\e^+$ and $\e^-$ beams are normally unpolarized, we have included the possibility of polarized beams, following the formalism of \cite{Ols80}. Thus the incoming $\e^+$ and $\e^-$ are characterized by polarizations $\mbf{P}^{\pm}$ in the rest frame of the particles: \begin{equation} \mbf{P}^{\pm} = P_{\mrm{T}}^{\pm} \hat{\mbf{s}}^{\pm} + P_{\mrm{L}}^{\pm} \hat{\mbf{p}}^{\pm} ~, \end{equation} where $0 \leq P_{\mrm{T}}^{\pm} \leq 1$ and $-1 \leq P_{\mrm{L}}^{\pm} \leq 1$, with the constraint \begin{equation} (\mbf{P}^{\pm})^2 = (P_{\mrm{T}}^{\pm})^2 + (P_{\mrm{L}}^{\pm})^2 \leq 1 ~. \end{equation} Here $\hat{\mbf{s}}^{\pm}$ are unit vectors perpendicular to the beam directions $\hat{\mbf{p}}^{\pm}$. To be specific, we choose a right-handed coordinate frame with $\hat{\mbf{p}}^{\pm} = (0,0, \mp 1)$, and standard transverse polarization directions (out of the machine plane for storage rings) $\hat{\mbf{s}}^{\pm} = (0, \pm 1,0)$, the latter corresponding to azimuthal angles $\varphi^{\pm} = \pm \pi /2$. As free parameters in the program we choose $P_{\mrm{L}}^+$, $P_{\mrm{L}}^-$, $P_{\mrm{T}} = \sqrt{P_{\mrm{T}}^+ P_{\mrm{T}}^-}$ and $\Delta \varphi = (\varphi^+ + \varphi^-) /2$. In the massless QED case, the probability to produce a flavour $\f$ is proportional to $e_{\f}^2$, i.e up-type quarks are four times as likely as down-type ones. In lowest-order massless QFD the corresponding relative probabilities are given by \cite{Ols80} \begin{eqnarray} h_{\f}(s) & = & e_{\e}^2 \, (1 - P_{\mrm{L}}^+ P_{\mrm{L}}^-) \, e_{\f}^2 \, + \, 2 e_{\e} \left\{ v_{\e} (1 - P_{\mrm{L}}^+ P_{\mrm{L}}^-) - a_{\e} (P_{\mrm{L}}^- - P_{\mrm{L}}^+) \right\} \, \Re\chi(s) \, e_{\f} v_{\f} \, + \nonumber \\ & & + \, \left\{ (v_{\e}^2 + a_{\e}^2) (1 - P_{\mrm{L}}^+ P_{\mrm{L}}^-) - 2 v_{\e} a_{\e} (P_{\mrm{L}}^- - P_{\mrm{L}}^+) \right\} \, \left| \chi(s) \right|^2 \, \left\{ v_{\f}^2 + a_{\f}^2 \right\} ~, \label{ee:hf} \end{eqnarray} where $\Re\chi(s)$ denotes the real part of $\chi(s)$. The $h_{\f}(s)$ expression depends both on the $s$ value and on the longitudinal polarization of the $\e^{\pm}$ beams in a non-trivial way. The cross section for the process $\ee \to \gammaZ \to \f \fbar$ may now be written as \begin{equation} \sigma_{\f}(s) = \frac{4 \pi \alphaem^2}{3 s} R_{\f}(s) ~, \end{equation} where $R_{\f}$ gives the ratio to the lowest-order QED cross section for the process $\ee \to \mu^+ \mu^-$, \begin{equation} R_{\f}(s) = N_C \, R_{\mrm{QCD}} \, h_{\f}(s) ~. \end{equation} The factor of $N_C = 3$ counts the number of colour states available for the $\q\qbar$ pair. The $R_{\mrm{QCD}}$ factor takes into account QCD loop corrections to the cross section. For $n_f$ effective flavours (normally $n_f =5$) \begin{equation} R_{\mrm{QCD}} \approx 1 + \frac{\alphas}{\pi} + (1.986 - 0.115 n_f) \left( \frac{\alphas}{\pi} \right)^2 + \cdots \label{ee:RQCD} \end{equation} in the $\br{\mrm{MS}}$ renormalization scheme \cite{Din79}. Note that $R_{\mrm{QCD}}$ does not affect the relative quark-flavour composition, and so is of peripheral interest in {\Je}. (For leptons the $N_C$ and $R_{\mrm{QCD}}$ factors would be absent, i.e. $N_C \, R_{\mrm{QCD}} = 1$, but leptonic final states are not generated in {\Je}.) Neglecting higher-order QCD and QFD effects, the corrections for massive quarks are given in terms of the velocity $v_{\q}$ of a quark with mass $m_{\q}$, $v_{\q} = \sqrt{ 1 - 4 m_{\q}^2 /s}$, as follows. The vector quark current terms in $h_{\f}$ (proportional to $e_{\f}^2$, $e_{\f} v_{\f}$, or $v_{\f}^2$) are multiplied by a threshold factor $v_{\q} (3 - v_{\q}^2) /2$, while the axial vector quark current term (proportional to $a_{\f}^2$) is multiplied by $v_{\q}^3$. While inclusion of quark masses in the QFD formulae decreases the total cross section, first-order QCD corrections tend in the opposite direction \cite{Jer81}. Na\"{\i}vely, one would expect one factor of $v_{\q}$ to get cancelled. So far, the available options are either to include threshold factors in full or not at all. Given that all five quarks are light at the scale of the $\Z^0$, the issue of quark masses is not really of interest at LEP. Here, however, purely weak corrections are important, in particular since they change the $\b$ quark partial width differently from that of the other ones \cite{Kuh89}. No such effects are included in the program. \subsubsection{First-order QCD matrix elements} The Born process $\ee \to \q \qbar$ is modified in first-order QCD by the probability for the $\q$ or $\qbar$ to radiate a gluon, i.e. by the process $\ee \to \q \qbar \g$. The matrix element is conveniently given in terms of scaled energy variables in the c.m. frame of the event, $x_1 = 2E_{\q}/E_{\mrm{cm}}$, $x_2 = 2E_{\qbar}/E_{\mrm{cm}}$, and $x_3 = 2E_{\g}/E_{\mrm{cm}}$, i.e. $x_1 + x_2 + x_3 = 2$. For massless quarks the matrix element reads \cite{Ell76} \begin{equation} \frac{1}{\sigma_0} \, \frac{\d \sigma}{\d x_1 \, \d x_2} = \frac{\alphas}{2\pi} \, C_F \, \frac{x_1^2 + x_2^2}{(1-x_1)(1-x_2)} ~, \label{ee:ME3j} \end{equation} where $\sigma_0$ is the lowest-order cross section, $C_F = 4/3$ is the appropriate colour factor, and the kinematically allowed region is $0 \leq x_i \leq 1, i = 1, 2, 3$. By kinematics, the $x_k$ variable for parton $k$ is related to the invariant mass $m_{ij}$ of the other two partons $i$ and $j$ by $y_{ij} = m_{ij}^2/E_{\mrm{cm}}^2 = 1 - x_k$. The strong coupling constant $\alphas$ is in first order given by \begin{equation} \alphas(Q^2) = \frac{12\pi}{(33-2n_f) \, \ln(Q^2/\Lambda^2)} ~. \label{ee:aS3j} \end{equation} Conventionally $Q^2 = s = E_{\mrm{cm}}^2$; we will return to this issue below. The number of flavours $n_f$ is 5 for LEP applications, and so the $\Lambda$ value determined is $\Lambda_5$ (while e.g. most deep inelastic scattering studies refer to $\Lambda_4$, the energies for these experiments being below the bottom threshold). The $\alphas$ values are matched at flavour thresholds, i.e. as $n_f$ is changed the $\Lambda$ value is also changed. It is therefore the derivative of $\alphas$ that changes at a threshold, not $\alphas$ itself. In order to separate 2-jets from 3-jets, it is useful to introduce jet-resolution parameters. This can be done in several different ways. Most famous are the $y$ and $(\epsilon, \delta)$ procedures. We will only refer to the $y$ cut, which is the one used in the program. Here a 3-parton configuration is called a 2-jet event if \begin{equation} \min_{i,j} (y_{ij}) = \min_{i,j} \left( \frac{m_{ij}^2}{E_{\mrm{cm}}^2} \right) < y ~. \end{equation} The cross section in eq. (\ref{ee:ME3j}) diverges for $x_1 \rightarrow 1$ or $x_2 \rightarrow 1$ but, when first-order propagator and vertex corrections are included, a corresponding singularity with opposite sign appears in the $\q \qbar$ cross section, so that the total cross section is finite. In analytical calculations, the average value of any well-behaved quantity ${\cal Q}$ can therefore be calculated as \begin{equation} \left\langle {\cal Q} \right\rangle = \frac{1}{\sigma_{\mrm{tot}}} \lim_{y \rightarrow 0} \left( {\cal Q}(\mrm{2parton}) \, \sigma_{\mrm{2parton}}(y) + \int_{y_{ij} > y} {\cal Q}(x_1,x_2) \, \frac{\d \sigma_{\mrm{3parton}}}{\d x_1 \, \d x_2} \, \d x_1 \, \d x_2 \right) ~, \label{ee:Obs} \end{equation} where any explicit $y$ dependence disappears in the limit $y \rightarrow 0$. In a Monte Carlo program, it is not possible to work with a negative total 2-jet rate, and thus it is necessary to introduce a fixed non-vanishing $y$ cut in the 3-jet phase space. Experimentally, there is evidence for the need of a low $y$ cut, i.e. a large 3-jet rate. For LEP applications, the recommended value is $y = 0.01$, which is about as far down as one can go and still retain a positive 2-jet rate. With $\alphas = 0.12$, in full second-order QCD (see below), the $2:3:4$ jet composition is then approximately $11 \% : 77 \% : 12 \%$. Note, however, that initial-state QED radiation may occasionally lower the c.m. energy significantly, i.e. increase $\alphas$, and thereby bring the 3-jet fraction above unity if $y$ is kept fixed at 0.01 also in those events. Therefore, at PETRA/PEP energies, $y$ values slightly above 0.01 are needed. In addition to the $y$ cut, the program contains a cut on the invariant mass $m_{ij}$ between any two partons, which is typically required to be larger than 2 GeV. This cut corresponds to the actual merging of two nearby parton jets, i.e. where a treatment with two separate partons rather than one would be superfluous in view of the smearing arising from the subsequent fragmentation. Since the cut-off mass scale $\sqrt{y} E_{\mrm{cm}}$ normally is much larger, this additional cut only enters for events at low energies. For massive quarks, the amount of QCD radiation is slightly reduced \cite{Iof78}: \begin{eqnarray} \frac{1}{\sigma_0} \, \frac{\d \sigma}{\d x_1 \, \d x_2} & = & \frac{\alphas}{2\pi} \, C_F \, \left\{ \frac{x_1^2 + x_2^2}{(1-x_1)(1-x_2)} - \frac{4 m_{\q}^2}{s} \left( \frac{1}{1-x_1} + \frac{1}{1-x_2} \right) \right. \nonumber \\[1mm] & & - \left. \frac{2 m_{\q}^2}{s} \left( \frac{1}{(1-x_1)^2} + \frac{1}{(1-x_2)^2} \right) - \frac{4 m_{\q}^4}{s^2} \left( \frac{1}{1-x_1} + \frac{1}{1-x_2} \right)^2 \right\} ~. \label{ee:threejMEmass} \end{eqnarray} In addition, the phase space for emission is reduced by the requirement \begin{equation} \frac{(1-x_1)(1-x_2)(1-x_3)}{x_3^2} \geq \frac{m_{\q}^2}{s} ~. \end{equation} For $\b$ quarks at LEP energies, these corrections are fairly small. \subsubsection{4-jet matrix elements} Two new event types are added in second-order QCD, $\ee \to \q \qbar \g \g$ and $\ee \to \q \qbar \q' \qbar'$. The 4-jet cross section has been calculated by several groups \cite{Ali80a,Gae80,Ell81,Dan82}, which agree on the result. The formulae are too lengthy to be quoted here. In one of the calculations \cite{Ali80a}, quark masses were explicitly included, but {\Je} only includes the massless expressions, as taken from \cite{Ell81}. Here the angular orientation of the event has been integrated out, so that five independent internal kinematical variables remain. These may be related to the six $y_{ij}$ and the four $y_{ijk}$ variables, $y_{ij} = m_{ij}^2 / s = (p_i + p_j)^2 / s$ and $y_{ijk} = m_{ijk}^2 / s = (p_i + p_j + p_k)^2 / s$, in terms of which the matrix elements are given. The original calculations were for the pure $\gamma$-exchange case; it was recently pointed out \cite{Kni89} that an additional contribution to the $\ee \to \q \qbar \q' \qbar'$ cross section arises from the axial part of the $\Z^0$. This term is not included in the program, but fortunately it is finite and small. Whereas the way the string, i.e. the fragmenting colour flux tube, is stretched is uniquely given in $\q \qbar \g$ event, for $\q \qbar \g \g$ events there are two possibilities: \mbox{$\q - \g_1 - \g_2 - \qbar$} or \mbox{$\q - \g_2 - \g_1 - \qbar$}. A knowledge of quark and gluon colours, obtained by perturbation theory, will uniquely specify the stretching of the string, as long as the two gluons do not have the same colour. The probability for the latter is down in magnitude by a factor $1 / N_C^2 = 1 / 9$. One may either choose to neglect these terms entirely, or to keep them for the choice of kinematical setup, but then drop them at the choice of string drawing \cite{Gus82}. We have adopted the latter procedure. Comparing the two possibilities, differences are typically 10--20\% for a given kinematical configuration, and less for the total 4-jet cross section, so from a practical point of view this is not a major problem. In higher orders, results depend on the renormalization scheme; we will use $\br{\mrm{MS}}$ throughout. In addition to this choice, several possible forms can be chosen for $\alphas$, all of which are equivalent to that order but differ in higher orders. We have picked the recommended standard \cite{PDG88} \begin{equation} \label{ee:aS4j} \alphas(Q^2) = \frac{12\pi}{(33-2n_f) \, \ln (Q^2 / \Lambda^2_{\br{\mrm{MS}}})} \left\{ 1 - 6 \, \frac{153-19n_f}{(33-2n_f)^2} \, \frac{\ln (\ln ( Q^2 / \Lambda^2_{\br{\mrm{MS}}}))} {\ln ( Q^2 / \Lambda^2_{\br{\mrm{MS}}})} \right\} ~. \end{equation} \subsubsection{Second-order 3-jet matrix elements} As for first order, a full second-order calculation consists both of real parton emission terms and of vertex and propagator corrections. These modify the 3-jet and 2-jet cross sections. Although there was some initial confusion, everybody soon agreed on the size of the loop corrections \cite{Ell81,Ver81,Fab82}. In analytic calculations, the procedure of eq. (\ref{ee:Obs}), suitably expanded, can therefore be used unambiguously for a well-behaved variable. For Monte Carlo event simulation, it is again necessary to impose some finite jet-resolution criterion. This means that four-parton events which fail the cuts should be reassigned either to the 3-jet or to the 2-jet event class. It is this area that caused quite a lot of confusion in the past \cite{Kun81,Got82,Ali82,Zhu83,Gut84,Gut87,Kra88}, and where full agreement does not exist. Most likely, agreement will never be reached, since there are indeed ambiguous points in the procedure, related to uncertainties on the theoretical side, as follows. For the $y$-cut case, any two partons with an invariant mass $m_{ij}^2 < y E_{\mrm{cm}}^2$ should be recombined into one. If the four-momenta are simply added, the sum will correspond to a parton with a positive mass, namely the original $m_{ij}$. The loop corrections are given in terms of final massless partons, however. In order to perform the (partial) cancellation between the four-parton real and the 3-parton virtual contributions, it is therefore necessary to get rid of the bothersome mass in the four-parton states. Several recombinations are used in practice, which go under names such as `E', `E0', `p' and `p0' \cite{OPA91}. In the `E'-type schemes, the energy of a recombined parton is given by $E_{ij} = E_i + E_j$, and three-momenta may have to be adjusted accordingly. In the `p'-type schemes, on the other hand, three-momenta are added, $\mbf{p}_{ij} = \mbf{p}_i + \mbf{p}_j$, and then energies may have to be adjusted. These procedures result in different 3-jet topologies, and therefore in different second-order differential 3-jet cross sections. Within each scheme, a number of lesser points remain to be dealt with, in particular what to do if a recombination of a nearby parton pair were to give an event with a non-$\q\qbar\g$ flavour structure. {\Je} contains two alternative second-order 3-jet implementations, GKS and ERT(Zhu). For historical reasons the former is default, but actually the latter is the recommended one today. Other parametrizations have also been made available that run together with {\Je}, see \cite{Sjo89,Mag89}. The GKS option is based on the GKS \cite{Gut84} calculation, where some of the original mistakes in FKSS \cite{Fab82} have been corrected. The GKS formulae have the advantage of giving the second-order corrections in closed analytic form, as not-too-long functions of $x_1$, $x_2$, and the $y$ cut. However, it is today recognized, also by the authors, that important terms are still missing, and that the matrix elements should therefore not be taken too seriously. The option is thus kept mainly for backwards compatibility. The ERT(Zhu) generator \cite{Zhu83} is based on the ERT matrix elements \cite{Ell81}, with a Monte Carlo recombination procedure suggested by Kunszt \cite{Kun81} and developed by Ali \cite{Ali82}. It has the merit of giving corrections in a convenient, parametrized form. For practical applications, the main limitation is that the corrections are only given for discrete values of the cut-off parameter $y$, namely $y$ = 0.01, 0.02, 0.03, 0.04, and 0.05. The basic approach is the following. Without any loss of generality, the full second-order 3-jet cross section can be written in terms of the `ratio function' $R(X,Y;y)$, defined by \begin{equation} \frac{1}{\sigma_0} \frac{\d \sigma_3^{\mrm{tot}}}{\d X \, \d Y} = \frac{\alphas}{\pi} A_0(X,Y) \left\{ 1 + \frac{\alphas}{\pi} R(X,Y;y) \right\} ~, \label{ee:Zhupar} \end{equation} where $X = x_1 - x_2 = x_{\q} - x_{\qbar}$, $Y = x_3 = x_g$, $\sigma_0$ is the lowest-order hadronic cross section, and $A_0(X,Y)$ the standard first-order 3-jet cross section, cf. eq. (\ref{ee:ME3j}). By Monte Carlo integration, the value of $R(X,Y;y)$ is evaluated in bins of $(X,Y)$, and the result parametrized by a simple function $F(X,Y;y)$. In order to obtain the second-order 3-jet rate, a small cut $y_0 = 10^{-7}$ was introduced. It was assumed that four-parton events which fail this cut can be (partly) cancelled analytically against the virtual 3-jet events, to give a net `regularized virtual' contribution to the 3-jet rate. For a given choice of $y$ cut, in the physical range $y \gg y_0$, an additional `soft' contribution comes from four-parton events which survive the $y_0$ cut but fail the $y$ one. A large sample (9\,000\,000) of four-parton events was generated inside the $y_0$ cut region. For events which failed the more stringent $y$ cuts, the parton pair with the smallest invariant mass was recombined into an effective jet, using the `p0' recombination scheme. This means that the individual three-momenta were added, $\mbf{p}_{ij} = \mbf{p}_i + \mbf{p}_j$, the mass of the recombined pair was set to zero for the calculation of energy, $E_{ij} = |\mbf{p}_i + \mbf{p}_j|$, and finally all four-momenta were rescaled by a common factor so as to preserve the correct c.m. frame energy. In calculating the ${\cal O}(\alphas^2)$ correction functions, care was taken to maintain the flavour signature of the jets in the recombination process. A quark and a gluon were recombined into a quark with the same flavour as the original quark, two gluons were recombined to form a gluon, etc. In some cases the three jets of the final state were not in the standard $\q \qbar \g$ configuration. The probability for this to happen corresponded to less than 0.5\% of the total cross section, even for the most stringent cuts used. For these non-$\q \qbar \g$ final states, the assignment of $\q$, $\qbar$ and $\g$ was done at random. The sum of `regularized virtual' (1\,000\,000 3-jet events were generated, with evaluated second-order weights) and `soft' corrections, normalized to the first-order 3-jet cross section, was tabulated in the $(X,Y)$ plane, using bins of size $0.05 \times 0.05$. This estimated $R$-function behaviour was then fit with a 12-parameter function $F$, \begin{eqnarray} F(X,Y;y)& = & p_1 + p_2 X^2 + p_3 X^4 + (p_4+p_5 X^2)Y+ (p_6+p_7X^2)Y^2 + \nonumber\\ & & (p_8+p_9X^2)Y^3 + p_{10}/(X^2-Y^2) + p_{11}/(1-Y) + p_{12}/Y ~. \label{ee:ZhuFun} \end{eqnarray} The parameters $p_i$ are reproduced in \cite{Sjo89}. \subsubsection{The matrix-element event generator scheme} The program contains parametrizations, separately, of the total first-order 3-jet rate, the total second-order 3-jet rate, and the total 4-jet rate, all as functions of $y$ (with $\alphas$ as a separate prefactor). These parametrizations have been obtained as follows: \begin{Itemize} \item The first-order 3-jet matrix element is almost analytically integrable; some small finite pieces were obtained by a truncated series expansion of the relevant integrand. \item The GKS second-order 3-jet matrix elements were integrated for 40 different $y$-cut values, evenly distributed in $\ln y$ between a smallest value $y = 0.001$ and the kinematical limit $y = 1/3$. For each $y$ value, 250\,000 phase-space points were generated, evenly in $\d \ln (1-x_i) = \d x_i/(1-x_i)$, $i = 1,2$, and the second-order 3-jet rate in the point evaluated. The properly normalized sum of weights in each of the 40 $y$ points were then fitted to a polynomial in $\ln(y^{-1}-2)$. For the ERT(Zhu) matrix elements the parametrizations in eq. (\ref{ee:ZhuFun}) were used to perform a corresponding Monte Carlo integration for the five $y$ values available. \item The 4-jet rate was integrated numerically, separately for $\q \qbar \g \g$ and $\q \qbar \q' \qbar'$ events, by generating large samples of 4-jet phase-space points within the boundary $y = 0.001$. Each point was classified according to the actual minimum $y$ between any two partons. The same events could then be used to update the summed weights for 40 different counters, corresponding to $y$ values evenly distributed in $\ln y$ between $y = 0.001$ and the kinematical limit $y = 1/6$. In fact, since the weight sums for large $y$ values only received contributions from few phase-space points, extra (smaller) subsamples of events were generated with larger $y$ cuts. The summed weights, properly normalized, were then parametrized in terms of polynomials in $\ln(y^{-1} - 5)$. Since it turned out to be difficult to obtain one single good fit over the whole range of $y$ values, different parametrizations are used above and below $y=0.018$. As originally given, the $\q \qbar \q' \qbar'$ parametrization only took into account four $\q'$ flavours, i.e. secondary $\b \bbar$ pairs were not generated, but this has been corrected for LEP. \end{Itemize} In the generation stage, each event is treated on its own, which means that the $\alphas$ and $y$ values may be allowed to vary from event to event. The main steps are the following. \begin{Enumerate} \item The $y$ value to be used in the current event is determined. If possible, this is the value given by you, but additional constraints exist from the validity of the parametrizations ($y \geq 0.001$ for GKS, $0.01 \leq y \leq 0.05$ for ERT(Zhu)) and an extra (user-modifiable) requirement of a minimum absolute invariant mass between jets (which translates into varying $y$ cuts due to the effects of initial-state QED radiation). \item The $\alphas$ value is calculated. \item For the $y$ and $\alphas$ values given, the relative two/three/4-jet composition is determined. This is achieved by using the parametrized functions of $y$ for 3- and 4-jet rates, multiplied by the relevant number of factors of $\alphas$. In ERT(Zhu), where the second-order 3-jet rate is available only at a few $y$ values, intermediate results are obtained by linear interpolation in the ratio of second-order to first-order 3-jet rates. The 3-jet and 4-jet rates are normalized to the analytically known second-order total event rate, i.e. divided by $R_{\mrm{QCD}}$ of eq. (\ref{ee:RQCD}). Finally, the 2-jet rate is obtained by conservation of total probability. \item If the combination of $y$ and $\alphas$ values is such that the total 3- plus 4-jet fraction is larger than unity, i.e. the remainder 2-jet fraction negative, the $y$-cut value is raised (for that event), and the process is started over at point 3. \item The choice is made between generating a 2-, 3- or 4-jet event, according to the relative probabilities. \item For the generation of 4-jets, it is first necessary to make a choice between $\q \qbar \g \g$ and $\q \qbar \q' \qbar'$ events, according to the relative (parametrized) total cross sections. A phase-space point is then selected, and the differential cross section at this point is evaluated and compared with a parametrized maximum weight. If the phase-space point is rejected, a new one is selected, until an acceptable 4-jet event is found. \item For 3-jets, a phase-space point is first chosen according to the first-order cross section. For this point, the weight \begin{equation} W(x_1,x_2;y) = 1 + \frac{\alphas}{\pi} R(x_1,x_2;y) \label{ee:WTJS} \end{equation} is evaluated. Here $R(x_1,x_2;y)$ is analytically given for GKS \cite{Gut84}, while it is approximated by the parametrization $F(X,Y;y)$ of eq. (\ref{ee:ZhuFun}) for ERT(Zhu). Again, linear interpolation of $F(X,Y;y)$ has to be applied for intermediate $y$ values. The weight $W$ is compared with a maximum weight \begin{equation} W_{\mmax}(y) = 1 + \frac{\alphas}{\pi} R_{\mmax}(y) ~, \end{equation} which has been numerically determined beforehand and suitably parametrized. If the phase-space point is rejected, a new point is generated, etc. \item Massive matrix elements are not available in {\Je} for second-order QCD (but are in the first-order option). However, if a 3- or 4-jet event determined above falls outside the phase-space region allowed for massive quarks, the event is rejected and reassigned to be a 2-jet event. (The way the $y_{ij}$ and $y_{ijk}$ variables of 4-jet events should be interpreted for massive quarks is not even unique, so some latitute has been taken here to provide a reasonable continuity from 3-jet events.) This procedure is known not to give the expected full mass suppression, but is a reasonable first approximation. \item Finally, if the event is classified as a 2-jet event, either because it was initially so assigned, or because it failed the massive phase-space cuts for 3- and 4-jets, the generation of 2-jets is trivial. \end{Enumerate} \subsubsection{Optimized perturbation theory} Theoretically, it turns out that the second-order corrections to the 3-jet rate are large. It is therefore not unreasonable to expect large third-order corrections to the 4-jet rate. Indeed, the experimental 4-jet rate is much larger than second order predicts (when fragmentation effects have been folded in), if $\alphas$ is determined based on the 3-jet rate \cite{Sjo84a,JAD88}. The only consistent way to resolve this issue is to go ahead and calculate the full next order. This is a tough task, however, so people have looked at possible shortcuts. For example, one can try to minimize the higher-order contributions by a suitable choice of the renormalization scale \cite{Ste81} --- `optimized perturbation theory'. This is equivalent to a different choice for the $Q^2$ scale in $\alphas$, a scale which is not unambiguous anyway. Indeed the standard value $Q^2 = s = E_{\mrm{cm}}^2$ is larger than the natural physical scale of gluon emission in events, given that most gluons are fairly soft. One could therefore pick another scale, $Q^2 = f s$, with $f < 1$. The ${\cal O}(\alphas)$ 3-jet rate would be increased by such a scale change, and so would the number of 4-jet events, including those which collapse into 3-jet ones. The loop corrections depend on the $Q^2$ scale, however, and compensate the changes above by giving a larger negative contribution to the 3-jet rate. The possibility of picking an optimized scale $f$ is implemented as follows \cite{Sjo89}. Assume that the differential 3-jet rate at scale $Q^2 = s$ is given by the expression \begin{equation} R_3 = r_1 \alphas + r_2 \alphas^2 ~, \end{equation} where $R_3$, $r_1$ and $r_2$ are functions of the kinematical variables $x_1$ and $x_2$ and the $y$ cut, as described above. When the coupling is chosen at a different scale, $Q'^2 = f s$, the 3-jet rate has to be changed to \begin{equation} R_3' = r_1' \alphas' + r_2 \alphas'^2 ~, \end{equation} where $r_1' = r_1$, \begin{equation} r_2' = r_2 + r_1 \frac{33-2n_f}{12\pi} \ln f ~, \label{ee:r2optim} \end{equation} and $\alphas' = \alphas(fs)$. Since we only have the Born term for 4-jets, here the effects of a scale change come only from the change in the coupling constant. Finally, the 2-jet cross section can still be calculated from the difference between the total cross section and the 3- and 4-jet cross sections. If an optimized scale is used in the program, the default value is $f=0.002$, which is favoured by the studies in ref. \cite{Bet89}. (In fact, it is also possible to use a correspondingly optimized $R_{\mrm{QCD}}$ factor, eq. (\ref{ee:RQCD}), but then the corresponding $f$ is chosen independently and much closer to unity.) The success of describing the jet rates should not hide the fact that one is dabbling in (educated, hopefully) guesswork, and that any conclusions based on this method have to be taken with a pinch of salt. One special problem associated with the use of optimized perturbation theory is that the differential 3-jet rate may become negative over large regions of the $(x_1, x_2)$ phase space. This problem already exists, at least in principle, even for a scale $f = 1$, since $r_2$ is not guaranteed to be positive definite. Indeed, depending on the choice of $y$ cut, $\alphas$ value and recombination scheme, one may observe a small region of negative differential 3-jet rate for the full second-order expression. This region is centred around $\q \qbar \g$ configurations, where the $\q$ and $\qbar$ are close together in one hemisphere and the $\g$ is alone in the other, i.e. $x_1 \approx x_2 \approx 1/2$. It is well understood why second-order corrections should be negative in this region \cite{Dok89}: the $\q$ and $\qbar$ of a $\q \qbar \g$ state are in a relative colour octet state, and thus the colour force between them is repulsive, which translates into a negative second-order term. However, as $f$ is decreased below unity, $r_2'$ receives a negative contribution from the $\ln f$ term, and the region of negative differential cross section has a tendency to become larger, also after taking into account related changes in $\alphas$. In an event-generator framework, where all events are supposed to come with unit weight, it is clearly not possible to simulate negative cross sections. What happens in the program is therefore that no 3-jet events at all are generated in the regions of negative differential cross section, and that the 3-jet rate in regions of positive cross sections is reduced by a constant factor, chosen so that the total number of 3-jet events comes out as it should. This is a consequence of the way the program works, where it is first decided what kind of event to generate, based on integrated 3-jet rates in which positive and negative contributions are added up with sign, and only thereafter the kinematics is chosen. Based on our physics understanding of the origin of this negative cross section, the approach adopted is as sensible as any, at least to that order in perturbation theory (what one might strive for is a properly exponentiated description of the relevant region). It can give rise to funny results for low $f$ values, however, as observed by OPAL \cite{OPA92} for the energy--energy correlation asymmetry. \subsubsection{Angular orientation} While pure $\gamma$ exchange gives a simple $1 + \cos^2\theta$ distribution for the $\q$ (and $\qbar$) direction in $\q \qbar$ events, $\Z^0$ exchange and $\gammaZ$ interference results in a forward--backward asymmetry. If one introduces \begin{eqnarray} h'_{\f}(s) & = & 2 e_{\e} \left\{ a_{\e} (1 - P_{\mrm{L}}^+ P_{\mrm{L}}^-) - v_{\e} (P_{\mrm{L}}^- - P_{\mrm{L}}^+) \right\} \, \Re\chi(s) e_{\f} a_{\f} \nonumber \\ & & + \, \left\{ 2 v_{\e} a_{\e} (1 - P_{\mrm{L}}^+ P_{\mrm{L}}^-) - (v_{\e}^2 + a_{\e}^2) (P_{\mrm{L}}^- - P_{\mrm{L}}^+) \right\} \, |\chi(s)|^2 \, v_{\f} a_{\f} ~, \end{eqnarray} then the angular distribution of the quark is given by \begin{equation} \frac{\d \sigma}{\d (\cos\theta_{\f})} \propto h_{\f}(s)(1 + \cos^2\theta_{\f}) + 2 h'_{\f}(s) \cos\theta_{\f} ~. \end{equation} The angular orientation of a 3- or 4-jet event may be described in terms of three angles $\chi$, $\theta$ and $\varphi$; for 2-jet events only $\theta$ and $\varphi$ are necessary. From a standard orientation, with the $\q$ along the $+z$ axis and the $\qbar$ in the $xz$ plane with $p_x > 0$, an arbitrary orientation may be reached by the rotations $+\chi$ in azimuthal angle, $+\theta$ in polar angle, and $+\varphi$ in azimuthal angle, in that order. Differential cross sections, including QFD effects and arbitrary beam polarizations have been given for 2- and 3-jet events in refs. \cite{Ols80,Sch80}. We use the formalism of ref. \cite{Ols80}, with $\chi \to \pi - \chi$ and $\varphi^- \to - (\varphi + \pi/2)$. The resulting formulae are tedious, but straightforward to apply, once the internal jet configuration has been chosen. 4-jet events are approximated by 3-jet ones, by joining the two gluons of a $\q \qbar \g \g$ event and the $\q'$ and $\qbar'$ of a $\q \qbar \q' \qbar'$ event into one effective jet. This means that some angular asymmetries are neglected \cite{Ali80a}, but that weak effects are automatically included. It is assumed that the second-order 3-jet events have the same angular orientation as the first-order ones, some studies on this issue may be found in \cite{Kor85}. Further, the formulae normally refer to the massless case; only for the QED 2- and 3-jet cases are mass corrections available. The main effect of the angular distribution of multijet events is to smear the lowest-order result, i.e. to reduce any anisotropies present in 2-jet systems. In the parton-shower option of the program, only the initial $\q \qbar$ axis is determined. The subsequent shower evolution then {\it de facto} leads to a smearing of the jet axis, although not necessarily in full agreement with the expectations from multijet matrix-element treatments. \subsubsection{Initial-state radiation} Initial-state photon radiation has been included using the formalism of ref. \cite{Ber82}. Here each event contains either no photon or one, i.e. it is a first-order non-exponentiated description. The main formula for the hard radiative photon cross section is \begin{equation} \frac{\d \sigma}{\d x_{\gamma}} = \frac{\alphaem}{\pi} \, \left( \ln\frac{s}{m_{\e}^2} -1 \right) \, \frac{1 + x_{\gamma}^2}{x_{\gamma}} \, \sigma_0 (\hat{s}) ~, \end{equation} where $x_{\gamma}$ is the photon energy fraction of the beam energy, $\hat{s} = (1-x_{\gamma}) s$ is the squared reduced hadronic c.m. energy, and $\sigma_0$ is the ordinary annihilation cross section at the reduced energy. In particular, the selection of jet flavours should be done according to expectations at the reduced energy. The cross section is divergent both for $x_{\gamma} \to 1$ and $x_{\gamma} \to 0$. The former is related to the fact that $\sigma_0$ has a $1/\hat{s}$ singularity (the real photon pole) for $\hat{s} \to 0$. An upper cut on $x_{\gamma}$ can here be chosen to fit the experimental setup. The latter is a soft photon singularity, which is to be compensated in the no-radiation cross section. A requirement $x_{\gamma} > 0.01$ has therefore been chosen so that the hard-photon fraction is smaller than unity. In the total cross section, effects from photons with $x_{\gamma} < 0.01$ are taken into account, together with vertex and vacuum polarization corrections (hadronic vacuum polarizations using a simple parametrization of the more complicated formulae of ref. \cite{Ber82}). The hard photon spectrum can be integrated analytically, for the full $\gammaZ$ structure including interference terms, provided that no new flavour thresholds are crossed and that the $R_{\mrm{QCD}}$ term in the cross section can be approximated by a constant over the range of allowed $\hat{s}$ values. In fact, threshold effects can be taken into account by standard rejection techniques, at the price of not obtaining the exact cross section analytically, but only by an effective Monte Carlo integration taking place in parallel with the ordinary event generation. In addition to $x_{\gamma}$, the polar angle $\theta_{\gamma}$ and azimuthal angle $\varphi_{\gamma}$ of the photons are also to be chosen. Further, for the orientation of the hadronic system, a choice has to be made whether the photon is to be considered as having been radiated from the $\e^+$ or from the $\e^-$. Final-state photon radiation, as well as interference between initial- and final-state radiation, has been left out of this treatment. The formulae for $\ee \to \mu^+ \mu^-$ cannot be simply taken over for the case of outgoing quarks, since the quarks as such only live for a short while before turning into hadrons. Another simplification in our treatment is that effects of incoming polarized $\e^{\pm}$ beams have been completely neglected, i.e. neither the effective shift in azimuthal distribution of photons nor the reduction in polarization is included. The polarization parameters of the program are to be thought of as the effective polarization surviving after initial-state radiation. \subsubsection{Alternative matrix elements} The program contains two sets of `toy model' matrix elements, one for an Abelian vector gluon model and one for a scalar gluon model. Clearly both of these alternatives are already excluded by data, and are anyway not viable alternatives for a consistent theory of strong interactions. They are therefore included more as references to show how well the characteristic features of QCD can be measured experimentally. Second-order matrix elements are available for the Abelian vector gluon model. These are easily obtained from the standard QCD matrix elements by a substitution of the Casimir group factors: $C_F = 4/3 \to 1$, $N_C = 3 \to 0$, and $T_R = n_{\f}/2 \to 3 n_{\f}$. First-order matrix elements contain only $C_F$; therefore the standard first-order QCD results may be recovered by a rescaling of $\alphas$ by a factor $4/3$. In second order the change of $N_C$ to 0 means that $\g \to \g\g$ couplings are absent from the Abelian model, while the change of $T_R$ corresponds to an enhancement of the $\g \to \q'\qbar'$ coupling, i.e. to an enhancement of the $\q\qbar\q'\qbar'$ 4-jet event rate. The second-order corrections to the 3-jet rate turn out to be strongly negative --- if $\alphas$ is fitted to get about the right rate of 4-jet events, the predicted differential 3-jet rate is negative almost everywhere in the $(x_1, x_2)$ plane. Whether this unphysical behaviour would be saved by higher orders is unclear. It has been pointed out that the rate can be made positive by a suitable choice of scale, since $\alphas$ runs in opposite directions in an Abelian model and in QCD \cite{Bet89}. This may be seen directly from eq. (\ref{ee:r2optim}), where the term $33 = 11 N_C$ is absent in the Abelian model, and therefore the scale-dependent term changes sign. In the program, optimized scales have not been implemented for this toy model. Therefore the alternatives provided for you are either to generate only 4-jet events, or to neglect second-order corrections to the 3-jet rate, or to have the total 3-jet rate set vanishing (so that only 2- and 4-jet events are generated). Normally we would expect the former to be the one of most interest, since it is in angular (and flavour) distributions of 4-jet events that the structure of QCD can be tested. Also note that the `correct' running of $\alphas$ is not included; you are expected to use the option where $\alphas$ is just given as a constant number. The scalar gluon model is even more excluded than the Abelian vector one, since differences appear already in the 3-jet matrix element \cite{Lae80}: \begin{equation} \frac{\d \sigma}{\d x_1 \, \d x_2} \propto \frac{x_3^2}{(1-x_1)(1-x_2)} \end{equation} when only $\gamma$ exchange is included. The axial part of the $\Z^0$ gives a slightly different shape; this is included in the program but does not make much difference. The angular orientation does include the full $\gammaZ$ interference \cite{Lae80}, but the main interest is in the 3-jet topology as such \cite{Ell79}. No higher-order corrections are included. It is recommended to use the option of a fixed $\alphas$ also here, since the correct running is not available. \subsection{Decays of Onia Resonances} \label{ss:oniadecays} Many different possibilities are open for the decay of heavy $J^{PC} = 1^{--}$ onia resonances. Of special interest are the decays into three gluons or two gluons plus a photon, since these offer unique possibilities to study a `pure sample' of gluon jets. A routine for this purpose is included in the program. It was written at a time where the expectations were to find toponium at PETRA energies. If, as now seems likely, the top mass is above 100~GeV, weak decays will dominate, to the extent that the top quark will decay weakly even before a bound toponium state is formed, and thus the routine will be of no use for top. The charm system, on the other hand, is far too low in mass for a jet language to be of any use. The only application is therefore likely to be for $\Upsilon$, which unfortunately also is on the low side in mass. The matrix element for $\q \qbar \to \g \g \g$ is (in lowest order) \cite{Kol78} \begin{equation} \frac{1}{\sigma_{\g \g \g}} \frac{\d \sigma_{\g \g \g}}{\d x_1 \, \d x_2} = \frac{1}{\pi^2 - 9} \left\{ \left( \frac{1-x_1}{x_2 x_3} \right)^2 + \left( \frac{1-x_2}{x_1 x_3} \right)^2 + \left( \frac{1-x_3}{x_1 x_2} \right)^2 \right\} ~, \label{ee:Upsilondec} \end{equation} where, as before, $x_i = 2 E_i / E_{\mrm{cm}}$ in the c.m. frame of the event. This is a well-defined expression, without the kind of singularities encountered in the $\q \qbar \g$ matrix elements. In principle, no cuts at all would be necessary, but for reasons of numerical simplicity we implement a $y$ cut as for continuum jet production, with all events not fulfilling this cut considered as (effective) $\g \g$ events. For $\g \g \g$ events, each $\g \g$ invariant mass is required to be at least 2 GeV. Another process is $\q \qbar \to \gamma \g \g$, obtained by replacing a gluon in $\q \qbar \to \g \g \g$ by a photon. This process has the same normalized cross section as the one above, if e.g. $x_1$ is taken to refer to the photon. The relative rate is \cite{Kol78} \begin{equation} \frac{\sigma_{\gamma \g \g}}{\sigma_{\g \g \g}} = \frac{36}{5} \, \frac{e_{\q}^2 \, \alphaem} {\alphas(Q^2)} ~. \end{equation} Here $e_{\q}$ is the charge of the heavy quark, and the scale in $\alphas$ has been chosen as the mass of the onium state. If the mass of the recoiling $\g \g$ system is lower than some cut-off (by default 2 GeV), the event is rejected. In the present implementation the angular orientation of the $\g \g \g$ and $\gamma \g \g$ events is given for the $\ee \to \gamma^* \to$ onium case \cite{Kol78} (optionally with beam polarization effects included), i.e. weak effects have not been included, since they are negligible at around 10~GeV. It is possible to start a perturbative shower evolution from either of the two states above. However, for $\Upsilon$ the phase space for additional evolution is so constrained that not much is to be gained from that. We therefore do not recommend this possibility. The shower generation machinery, when starting up from a $\gamma \g \g$ configuration, is constructed such that the photon energy is not changed. This means that there is currently no possibility to use showers to bring the theoretical photon spectrum in better agreement with the experimental one. In string fragmentation language, a $\g \g \g$ state corresponds to a closed string triangle with the three gluons at the corners. As the partons move apart from a common origin, the string triangle expands. Since the photon does not take part in the fragmentation, the $\gamma \g \g$ state corresponds to a double string running between the two gluons. \subsection{Routines and Common Block Variables} \label{ss:eeroutines} \subsubsection{$\ee$ continuum event generation} The only routine a normal user will call to generate $\ee$ continuum events is {\tt LUEEVT}. The other routines listed below, as well as {\tt LUSHOW} (see section \ref{ss:showrout}), are called by {\tt LUEEVT}. \drawbox{CALL LUEEVT(KFL,ECM)}\label{p:LUEEVT} \begin{entry} \itemc{Purpose:} to generate a complete event $\ee \to \gammaZ \to \q\qbar \to$ parton shower $\to$ hadrons according to QFD and QCD cross sections. As an alternative to parton showers, second-order matrix elements are available for $\q\qbar + \q\qbar\g + \q\qbar\g\g + \q\qbar\q'\qbar'$ production. \iteme{KFL :} flavour of events generated. \begin{subentry} \iteme{= 0 :} mixture of all allowed flavours according to relevant probabilities. \iteme{= 1 - 8 :} primary quarks are only of the specified flavour {\tt KFL}. \end{subentry} \iteme{ECM :} total c.m. energy of system. \itemc{Remark:} Each call generates one event, which is independent of preceding ones, with one exception, as follows. If radiative corrections are included, the shape of the hard photon spectrum is recalculated only with each {\tt LUXTOT} call, which normally is done only if {\tt KFL}, {\tt ECM} or {\tt MSTJ(102)} is changed. A change of e.g. the $\Z^0$ mass in mid-run has to be followed either by a user call to {\tt LUXTOT} or by an internal call forced e.g. by putting {\tt MSTJ(116)=3}. \end{entry} \boxsep \begin{entry} \iteme{SUBROUTINE LUXTOT(KFL,ECM,XTOT) :}\label{p:LUXTOT} to calculate the total hadronic cross section, including quark thresholds, weak, beam polarization, and QCD effects and radiative corrections. In the process, variables necessary for the treatment of hard photon radiation are calculated and stored. \begin{subentry} \iteme{KFL, ECM :} as for {\tt LUEEVT}. \iteme{XTOT :} the calculated total cross section in nb. \end{subentry} \iteme{SUBROUTINE LURADK(ECM,MK,PAK,THEK,PHIK,ALPK) :}\label{p:LURADK} to describe initial-state hard $\gamma$ radiation. \iteme{SUBROUTINE LUXKFL(KFL,ECM,ECMC,KFLC) :}\label{p:LUXKFL} to generate the primary quark flavour in case this is not specified by the user. \iteme{SUBROUTINE LUXJET(ECM,NJET,CUT) :}\label{p:LUXJET} to determine the number of jets (2, 3 or 4) to be generated within the kinematically allowed region (characterized by {\tt CUT} $= y_{\mrm{cut}}$) in the matrix-element approach; to be chosen such that all probabilities are between 0 and 1. \iteme{SUBROUTINE LUX3JT(NJET,CUT,KFL,ECM,X1,X2) :}\label{p:LUX3JT} to generate the internal momentum variables of a 3-jet event, $\q\qbar\g$, according to first- or second-order QCD matrix elements. \iteme{SUBROUTINE LUX4JT(NJET,CUT,KFL,ECM,KFLN,X1,X2,X4,X12,X14) :}% \label{p:LUX4JT} to generate the internal momentum variables for a 4-jet event, $\q\qbar\g\g$ or $\q\qbar\q'\qbar'$, according to second-order QCD matrix elements. \iteme{SUBROUTINE LUXDIF(NC,NJET,KFL,ECM,CHI,THE,PHI) :}\label{p:LUXDIF} to describe the angular orientation of the jets. In first-order QCD the complete QED or QFD formulae are used; in second order 3-jets are assumed to have the same orientation as in first, and 4-jets are approximated by 3-jets. \end{entry} \subsubsection{A routine for onium decay} In {\tt LUONIA} we have implemented the decays of heavy onia resonances into three gluons or two gluons plus a photon, which are the dominant non-background-like decays of $\Upsilon$. \drawbox{CALL LUONIA(KFL,ECM)}\label{p:LUONIA} \begin{entry} \itemc{Purpose:} to simulate the process $\ee \to \gamma^* \to 1^{--}$ onium resonance $\to (\g\g\g$ or $\g\g\gamma) \to$ shower $\to$ hadrons. \iteme{KFL :} the flavour of the quark giving rise to the resonance. \begin{subentry} \iteme{= 0 :} generate $\g\g\g$ events alone. \iteme{= 1 - 8 :} generate $\g\g\g$ and $\g\g\gamma$ events in mixture determined by the squared charge of flavour {\tt KFL}. Normally {\tt KFL=} 5 or 6. \end{subentry} \iteme{ECM :} total c.m. energy of system. \end{entry} \subsubsection{Common block variables} The status codes and parameters relevant for the $\ee$ routines are found in the common block {\tt LUDAT1}. This common block also contains more general status codes and parameters, described elsewhere. \drawbox{COMMON/LUDAT1/MSTU(200),PARU(200),MSTJ(200),PARJ(200)} \begin{entry} \itemc{Purpose:} to give access to a number of status codes and parameters regulating the performance of the $\ee$ event generation routines. \iteme{MSTJ(101) :}\label{p:MSTJ101} (D=5) gives the type of QCD corrections used for continuum events. \begin{subentry} \iteme{= 0 :} only $\q\qbar$ events are generated. \iteme{= 1 :} $\q\qbar + \q\qbar\g$ events are generated according to first-order QCD. \iteme{= 2 :} $\q\qbar + \q\qbar\g + \q\qbar\g\g + \q\qbar\q'\qbar'$ events are generated according to second-order QCD. \iteme{= 3 :} $\q\qbar + \q\qbar\g + \q\qbar\g\g + \q\qbar\q'\qbar'$ events are generated, but without second-order corrections to the 3-jet rate. \iteme{= 5 :} a parton shower is allowed to develop from an original $\q\qbar$ pair, see {\tt MSTJ(41) - MSTJ(50)} for details. \iteme{= -1 :} only $\q\qbar\g$ events are generated (within same matrix-element cuts as for {\tt =1}). Since the change in flavour composition from mass cuts or radiative corrections is not taken into account, this option is not intended for quantitative studies. \iteme{= -2 :} only $\q\qbar\g\g$ and $\q\qbar\q'\qbar'$ events are generated (as for {\tt =2}). The same warning as for {\tt =-1} applies. \iteme{= -3 :} only $\q\qbar\g\g$ events are generated (as for {\tt =2}). The same warning as for {\tt =-1} applies. \iteme{= -4 :} only $\q\qbar\q'\qbar'$ events are generated (as for {\tt =2}). The same warning as for {\tt =-1} applies. \itemc{Note 1:} {\tt MSTJ(101)} is also used in {\tt LUONIA}, with \iteme{$\leq$ 4 :} $\g\g\g + \gamma\g\g$ events are generated according to lowest-order matrix elements. \iteme{$\geq$ 5 :} a parton shower is allowed to develop from the original $\g\g\g$ or $\g\g\gamma$ configuration, see {\tt MSTJ(41) - MSTJ(50)} for details. \itemc{Note 2:} The default values of fragmentation parameters have been chosen to work well with the default parton-shower approach above. If any of the other options are used, or if the parton shower is used in non-default mode, it may be necessary to retune fragmentation parameters. As an example, we note that the second-order matrix-element approach ({\tt MSTJ(101)=2}) at PETRA/PEP energies gives a better description when the $a$ and $b$ parameters of the symmetric fragmentation function are set to $a =${\tt PARJ(41)=1}, $b =${\tt PARJ(42)=0.7}, and the width of the transverse momentum distribution to $\sigma =${\tt PARJ(21)=0.40}. In principle, one also ought to change the joining parameter to {\tt PARJ(33)=PARJ(35)=1.1} to preserve a flat rapidity plateau, but if this should be forgotten, it does not make too much difference. For applications at TRISTAN or LEP, one must expect to have to change the matrix-element approach parameters even more, to make up for additional soft gluon effects not covered in this approach. \end{subentry} \iteme{MSTJ(102) :} (D=2) inclusion of weak effects ($\Z^0$ exchange) for flavour production, angular orientation, cross sections and initial-state photon radiation in continuum events. \begin{subentry} \iteme{= 1 :} QED, i.e. no weak effects are included. \iteme{= 2 :} QFD, i.e. including weak effects. \iteme{= 3 :} as {\tt =2}, but at initialization in {\tt LUXTOT} the $\Z^0$ width is calculated from $\ssintw$, $\alphaem$ and $\Z^0$ and quark masses (including bottom and top threshold factors for {\tt MSTJ(103)} odd), assuming three full generations, and the result is stored in {\tt PARJ(124)}. \end{subentry} \iteme{MSTJ(103) :} (D=7) mass effects in continuum matrix elements, in the form {\tt MSTJ(103)} $= M_1 + 2M_2 + 4M_3$, where $M_i = 0$ if no mass effects and $M_i = 1$ if mass effects should be included. Here; \begin{subentry} \iteme{$M_1$ :} threshold factor for new flavour production according to QFD result; \iteme{$M_2$ :} gluon emission probability (only applies for {\tt |MSTJ(101)|}$\leq 1$, otherwise no mass effects anyhow); \iteme{$M_3$ :} angular orientation of event (only applies for {\tt |MSTJ(101)|}$\leq 1$ and {\tt MSTJ(102)=1}, otherwise no mass effects anyhow). \end{subentry} \iteme{MSTJ(104) :} (D=5) number of allowed flavours, i.e. flavours that can be produced in a continuum event if the energy is enough. A change to 6 makes top production allowed above the threshold, etc. Note that in $\q\qbar\q'\qbar'$ events only the first five flavours are allowed in the secondary pair, produced by a gluon breakup. \iteme{MSTJ(105) :} (D=1) fragmentation and decay in {\tt LUEEVT} and {\tt LUONIA} calls. \begin{subentry} \iteme{= 0 :} no {\tt LUEXEC} calls, i.e. only matrix-element and/or parton-shower treatment. \iteme{= 1 :} {\tt LUEXEC} calls are made to generate fragmentation and decay chain. \iteme{= -1 :} no {\tt LUEXEC} calls and no collapse of small jet systems into one or two particles (in {\tt LUPREP}). \end{subentry} \iteme{MSTJ(106) :} (D=1) angular orientation in {\tt LUEEVT} and {\tt LUONIA}. \begin{subentry} \iteme{= 0 :} standard orientation of events, i.e. $\q$ along $+z$ axis and $\qbar$ along $-z$ axis or in $xz$ plane with $p_x > 0$ for continuum events, and $\g_1\g_2\g_3$ or $\gamma\g_2\g_3$ in $xz$ plane with $\g_1$ or $\gamma$ along the $+z$ axis for onium events. \iteme{= 1 :} random orientation according to matrix elements. \end{subentry} \iteme{MSTJ(107) :} (D=0) radiative corrections to continuum events. \begin{subentry} \iteme{= 0 :} no radiative corrections. \iteme{= 1 :} initial-state radiative corrections (including weak effects for {\tt MSTJ(102)=} 2 or 3). \end{subentry} \iteme{MSTJ(108) :} (D=2) calculation of $\alphas$ for matrix-element alternatives. The {\tt MSTU(111)} and {\tt PARU(112)} values are automatically overwritten in {\tt LUEEVT} or {\tt LUONIA} calls accordingly. \begin{subentry} \iteme{= 0 :} fixed $\alphas$ value as given in {\tt PARU(111)}. \iteme{= 1 :} first-order formula is always used, with $\Lambda_{\mrm{QCD}}$ given by {\tt PARJ(121)}. \iteme{= 2 :} first- or second-order formula is used, depending on value of {\tt MSTJ(101)}, with $\Lambda_{\mrm{QCD}}$ given by {\tt PARJ(121)} or {\tt PARJ(122)}. \end{subentry} \iteme{MSTJ(109) :} (D=0) gives a possibility to switch from QCD matrix elements to some alternative toy models. Is not relevant for shower evolution, {\tt MSTJ(101)=5}, where one can use {\tt MSTJ(49)} instead. \begin{subentry} \iteme{= 0 :} standard QCD scenario. \iteme{= 1 :} a scalar gluon model. Since no second-order corrections are available in this scenario, one can only use this with {\tt MSTJ(101) = 1} or {\tt -1}. Also note that the event-as-a-whole angular distribution is for photon exchange only (i.e. no weak effects), and that no higher-order corrections to the total cross section are included. \iteme{= 2 :} an Abelian vector gluon theory, with the colour factors $C_F = 1$ ($= 4/3$ in QCD), $N_C = 0$ ($= 3$ in QCD) and $T_R = 3 n_f$ ($= n_f/2$ in QCD). If one selects $\alpha_{\mrm{Abelian}} = (4/3) \alpha_{\mrm{QCD}}$, the 3-jet cross section will agree with the QCD one, and differences are to be found only in 4-jets. The {\tt MSTJ(109)=2} option has to be run with {\tt MSTJ(110)=1} and {\tt MSTJ(111)=0}; if need be, the latter variables will be overwritten by the program. \\ {\bf Warning:} second-order corrections give a large negative contribution to the 3-jet cross section, so large that the whole scenario is of doubtful use. In order to make the second-order options work at all, the 3-jet cross section is here by hand set exactly equal to zero for {\tt MSTJ(101)=2}. It is here probably better to use the option {\tt MSTJ(101)=3}, although this is not a consistent procedure either. \end{subentry} \iteme{MSTJ(110) :} (D=2) choice of second-order contributions to the 3-jet rate. \begin{subentry} \iteme{= 1 :} the GKS second-order matrix elements, i.e. the old {\Je} standard. \iteme{= 2 :} the Zhu parametrization of the ERT matrix elements, based on the program of Kunszt and Ali, i.e. in historical sequence ERT/Kunszt/Ali/Zhu. The parametrization is available for $y =$ 0.01, 0.02, 0.03, 0.04 and 0.05. Values outside this range are put at the nearest border, while those inside it are given by a linear interpolation between the two nearest points. Since this procedure is rather primitive, one should try to work at one of the values given above. Note that no Abelian QCD parametrization is available for this option. \end{subentry} \iteme{MSTJ(111) :} (D=0) use of optimized perturbation theory for second-order matrix elements (it can also be used for first-order matrix elements, but here it only corresponds to a trivial rescaling of the $\alphas$ argument). \begin{subentry} \iteme{= 0 :} no optimization procedure; i.e. $Q^2 = E_{\mrm{cm}}^2$. \iteme{= 1 :} an optimized $Q^2$ scale is chosen as $Q^2 = f E_{\mrm{cm}}^2$, where $f =${\tt PARJ(128)} for the total cross section $R$ factor, while $f =${\tt PARJ(129)} for the 3- and 4-jet rates. This $f$ value enters via the $\alphas$, and also via a term proportional to $\alphas^2 \ln f$. Some constraints are imposed; thus the optimized `3-jet' contribution to $R$ is assumed to be positive (for {\tt PARJ(128)}), the total 3-jet rate is not allowed to be negative (for {\tt PARJ(129)}), etc. However, there is no guarantee that the differential 3-jet cross section is not negative (and truncated to 0) somewhere (this can also happen with $f = 1$, but is then less frequent). The actually obtained $f$ values are stored in {\tt PARJ(168)} and {\tt PARJ(169)}, respectively. If an optimized $Q^2$ scale is used, then the $\Lambda_{\mrm{QCD}}$ (and $\alphas$) should also be changed. With the value $f = 0.002$, it has been shown \cite{Bet89} that a $\Lambda_{\mrm{QCD}} = 0.100$ GeV gives a reasonable agreement; the parameter to be changed is {\tt PARJ(122)} for a second-order running $\alphas$. Note that, since the optimized $Q^2$ scale is sometimes below the charm threshold, the effective number of flavours used in $\alphas$ may well be 4 only. If one feels that it is still appropriate to use 5 flavours (one choice might be as good as the other), it is necessary to put {\tt MSTU(113)=5}. \end{subentry} \iteme{MSTJ(115) :} (D=1) documentation of continuum or onium events, in increasing order of completeness. \begin{subentry} \iteme{= 0 :} only the parton shower, the fragmenting partons and the generated hadronic system are stored in the {\tt LUJETS} common block. \iteme{= 1 :} also a radiative photon is stored (for continuum events). \iteme{= 2 :} also the original $\ee$ are stored (with {\tt K(I,1)=21}). \iteme{= 3 :} also the $\gamma$ or $\gammaZ$ exchanged for continuum events, the onium state for resonance events is stored (with {\tt K(I,1)=21}). \end{subentry} \iteme{MSTJ(116) :} (D=1) initialization of total cross section and radiative photon spectrum in {\tt LUEEVT} calls. \begin{subentry} \iteme{= 0 :} never; cannot be used together with radiative corrections. \iteme{= 1 :} calculated at first call and then whenever {\tt KFL} or {\tt MSTJ(102)} is changed or {\tt ECM} is changed by more than {\tt PARJ(139)}. \iteme{= 2 :} calculated at each call. \iteme{= 3 :} everything is reinitialized in the next call, but {\tt MSTJ(116)} is afterwards automatically put {\tt =1} for use in subsequent calls. \end{subentry} \iteme{MSTJ(119) :} (I) check on need to reinitialize {\tt LUXTOT}. \iteme{MSTJ(120) :} (R) type of continuum event generated with the matrix-element option (with the shower one, the result is always {\tt =1}). \begin{subentry} \iteme{= 1 :} $\q\qbar$. \iteme{= 2 :} $\q\qbar\g$. \iteme{= 3 :} $\q\qbar\g\g$ from Abelian (QED-like) graphs in matrix element. \iteme{= 4 :} $\q\qbar\g\g$ from non-Abelian (i.e. containing triple-gluon coupling) graphs in matrix element. \iteme{= 5 :} $\q\qbar\q'\qbar'$. \end{subentry} \iteme{MSTJ(121) :} (R) flag set if a negative differential cross section was encountered in the latest {\tt LUX3JT} call. Events are still generated, but maybe not quite according to the distribution one would like (the rate is set to zero in the regions of negative cross section, and the differential rate in the regions of positive cross section is rescaled to give the `correct' total 3-jet rate). \boxsep \iteme{PARJ(121) :}\label{p:PARJ121} (D=1.0 GeV) $\Lambda$ value used in first-order calculation of $\alphas$ in the matrix-element alternative. \iteme{PARJ(122) :} (D=0.25 GeV) $\Lambda$ values used in second-order calculation of $\alphas$ in the matrix-element alternative. \iteme{PARJ(123) :} (D=91.187 GeV) mass of $\Z^0$ as used in propagators for the QFD case. \iteme{PARJ(124) :} (D=2.489 GeV) width of $\Z^0$ as used in propagators for the QFD case. Overwritten at initialization if {\tt MSTJ(102)=3}. \iteme{PARJ(125) :} (D=0.01) $y_{\mrm{cut}}$, minimum squared scaled invariant mass of any two partons in 3- or 4-jet events; the main user-controlled matrix-element cut. {\tt PARJ(126)} provides an additional constraint. For each new event, it is additionally checked that the total 3- plus 4-jet fraction does not exceed unity; if so the effective $y$ cut will be dynamically increased. The actual $y$-cut value is stored in {\tt PARJ(150)}, event by event. \iteme{PARJ(126) :} (D=2. GeV) minimum invariant mass of any two partons in 3- or 4-jet events; a cut in addition to the one above, mainly for the case of a radiative photon lowering the hadronic c.m. energy significantly. \iteme{PARJ(127) :} (D=1. GeV) is used as a safety margin for small colour-singlet jet systems, cf. {\tt PARJ(32)}, specifically $\q\qbar'$ masses in $\q\qbar\q'\qbar'$ 4-jet events and $\g\g$ mass in onium $\gamma\g\g$ events. \iteme{PARJ(128) :} (D=0.25) optimized $Q^2$ scale for the QCD $R$ (total rate) factor for the {\tt MSTJ(111)=1} option is given by $Q^2 = f E_{\mrm{cm}}^2$, where $f =${\tt PARJ(128)}. For various reasons the actually used $f$ value may be increased compared with the nominal one; while {\tt PARJ(128)} gives the nominal value, {\tt PARJ(168)} gives the actual one for the current event. \iteme{PARJ(129) :} (D=0.002) optimized $Q^2$ scale for the 3- and 4-jet rate for the {\tt MSTJ(111)=1} option is given by $Q^2 = f E_{\mrm{cm}}^2$, where $f =${\tt PARJ(129)}. For various reasons the actually used $f$ value may be increased compared with the nominal one; while {\tt PARJ(129)} gives the nominal value, {\tt PARJ(169)} gives the actual one for the current event. The default value is in agreement with the studies of Bethke \cite{Bet89}. \iteme{PARJ(131), PARJ(132) :} (D=2*0.) longitudinal polarizations $P_{\mrm{L}}^+$ and $P_{\mrm{L}}^-$ of incoming $\e^+$ and $\e^-$. \iteme{PARJ(133) :} (D=0.) transverse polarization $P_{\mrm{T}} = \sqrt{P_{\mrm{T}}^+ P_{\mrm{T}}^-}$, with $P_{\mrm{T}}^+$ and $P_{\mrm{T}}^-$ transverse polarizations of incoming $\e^+$ and $\e^-$. \iteme{PARJ(134) :} (D=0.) mean of transverse polarization directions of incoming $\e^+$ and $\e^-$, $\Delta \varphi = (\varphi^+ + \varphi^-) /2$, with $\varphi$ the azimuthal angle of polarization, leading to a shift in the $\varphi$ distribution of jets by $\Delta \varphi$. \iteme{PARJ(135) :} (D=0.01) minimum photon energy fraction (of beam energy) in initial-state radiation; should normally never be changed (if lowered too much, the fraction of events containing a radiative photon will exceed unity, leading to problems). \iteme{PARJ(136) :} (D=0.99) maximum photon energy fraction (of beam energy) in initial-state radiation; may be changed to reflect actual trigger conditions of a detector (but must always be larger than {\tt PARJ(135)}). \iteme{PARJ(139) :} (D=0.2 GeV) maximum deviation of $E_{\mrm{cm}}$ from the corresponding value at last {\tt LUXTOT} call, above which a new call is made if {\tt MSTJ(116)=1}. \iteme{PARJ(141) :} (R) value of $R$, the ratio of continuum cross section to the lowest-order muon pair production cross section, as given in massless QED (i.e. three times the sum of active quark squared charges, possibly modified for polarization). \iteme{PARJ(142) :} (R) value of $R$ including quark-mass effects (for {\tt MSTJ(102)=1}) and/or weak propagator effects (for {\tt MSTJ(102)=2}). \iteme{PARJ(143) :} (R) value of $R$ as {\tt PARJ(142)}, but including QCD corrections as given by {\tt MSTJ(101)}. \iteme{PARJ(144) :} (R) value of $R$ as {\tt PARJ(143)}, but additionally including corrections from initial-state photon radiation (if {\tt MSTJ(107)=1}). Since the effects of heavy flavour thresholds are not simply integrable, the initial value of {\tt PARJ(144)} is updated during the course of the run to improve accuracy. \iteme{PARJ(145) - PARJ(148) :} (R) absolute cross sections in nb as for the cases {\tt PARJ(141) - PARJ(144)} above. \iteme{PARJ(150) :} (R) current effective matrix element cut-off $y_{\mrm{cut}}$, as given by {\tt PARJ(125), PARJ(126)} and the requirements of having non-negative cross sections for 2-, 3- and 4-jet events. Not used in parton showers. \iteme{PARJ(151) :} (R) value of c.m. energy {\tt ECM} at last {\tt LUXTOT} call. \iteme{PARJ(152) :} (R) current first-order contribution to the 3-jet fraction; modified by mass effects. Not used in parton showers. \iteme{PARJ(153) :} (R) current second-order contribution to the 3-jet fraction; modified by mass effects. Not used in parton showers. \iteme{PARJ(154) :} (R) current second-order contribution to the 4-jet fraction; modified by mass effects. Not used in parton showers. \iteme{PARJ(155) :} (R) current fraction of 4-jet rate attributable to $\q\qbar\q'\qbar'$ events rather than $\q\qbar\g\g$ ones; modified by mass effects. Not used in parton showers. \iteme{PARJ(156) :} (R) has two functions when using second-order QCD. For a 3-jet event, it gives the ratio of the second-order to the total 3-jet cross section in the given kinematical point. For a 4-jet event, it gives the ratio of the modified 4-jet cross section, obtained when neglecting interference terms whose colour flow is not well defined, to the full unmodified one, all evaluated in the given kinematical point. Not used in parton showers. \iteme{PARJ(157) - PARJ(159) :} (I) used for cross-section calculations to include mass threshold effects to radiative photon cross section. What is stored is basic cross section, number of events generated and number that passed cuts. \iteme{PARJ(160) :} (R) nominal fraction of events that should contain a radiative photon. \iteme{PARJ(161) - PARJ(164) :} (I) give shape of radiative photon spectrum including weak effects. \iteme{PARJ(168) :} (R) actual $f$ value of current event in optimized perturbation theory for $R$; see {\tt MSTJ(111)} and {\tt PARJ(128)}. \iteme{PARJ(169) :} (R) actual $f$ value of current event in optimized perturbation theory for 3- and 4-jet rate; see {\tt MSTJ(111)} and {\tt PARJ(129)}. \iteme{PARJ(171) :} (R) fraction of cross section corresponding to the axial coupling of quark pair to the intermediate $\gammaZ$ state; needed for the Abelian gluon model 3-jet matrix element. \end{entry} \subsection{Examples} An ordinary $\ee$ annihilation event in the continuum, at a c.m. energy of 40 GeV, may be generated with \begin{verbatim} CALL LUEEVT(0,40.) \end{verbatim} In this case a $\q\qbar$ event is generated, including weak effects, followed by parton-shower evolution and fragmentation/decay treatment. Before a call to {\tt LUEEVT}, however, a number of default values may be changed, e.g. {\tt MSTJ(101)=2} to use second-order QCD matrix elements, giving a mixture of $\q\qbar$, $\q\qbar\g$, $\q\qbar\g\g$, and $\q\qbar\q'\qbar'$ events, {\tt MSTJ(102)=1} to have QED only, {\tt MSTJ(104)=6} to allow $\t\tbar$ production as well, {\tt MSTJ(107)=1} to include initial-state photon radiation (including a treatment of the $\Z^0$ pole), {\tt PARJ(123)=92.0} to change the $\Z^0$ mass, {\tt PARJ(81)=0.3} to change the parton-shower $\Lambda$ value, or {\tt PARJ(82)=1.5} to change the parton-shower cut-off. If initial-state photon radiation is used, some restrictions apply to how one can alternate the generation of events at different energies or with different $\Z^0$ mass, etc. These restrictions are not there for efficiency reasons (the extra time for recalculating the extra constants every time is small), but because it ties in with the cross-section calculations (see {\tt PARJ(144)}). Most parameters can be changed independently of each other. However, if just one or a few parameters/switches are changed, one should not be surprised to find a rather bad agreement with the data, like e.g. a too low or high average hadron multiplicity. It is therefore usually necessary to retune one parameter related to the perturbative QCD description, like $\alphas$ or $\Lambda$, one of the two parameters $a$ and $b$ of the Lund symmetric fragmentation function (since they are so strongly correlated, it is often not necessary to retune both of them), and the average fragmentation transverse momentum --- see Note~2 of the {\tt MSTJ(101)} description for an example. For very detailed studies it may be necessary to retune even more parameters. The three-gluon and gluon--gluon--photon decays of $\Upsilon$ may be simulated by a call \begin{verbatim} CALL LUONIA(5,9.46) \end{verbatim} Unfortunately, with present top-mass limits, this routine will not be of much interest for toponium studies (weak decays will dominate). A typical program for analysis of $\ee$ annihilation events at 100 GeV might look something like \begin{verbatim} COMMON/LUJETS/N,K(4000,5),P(4000,5),V(4000,5) COMMON/LUDAT1/MSTU(200),PARU(200),MSTJ(200),PARJ(200) COMMON/LUDAT2/KCHG(500,3),PMAS(500,4),PARF(2000),VCKM(4,4) COMMON/LUDAT3/MDCY(500,3),MDME(2000,2),BRAT(2000),KFDP(2000,5) MDCY(LUCOMP(111),1)=0 ! put pi0 stable MSTJ(107)=1 ! include initial-state radiation PARU(41)=1. ! use linear sphericity ..... ! other desired changes ..... ! initialize analysis statistics DO 100 IEVENT=1,1000 ! loop over events CALL LUEEVT(0,100.) ! generate new event IF(IEVENT.EQ.1) CALL LULIST(2) ! list first event CALL LUTABU(11) ! save particle composition ! statistics CALL LUEDIT(2) ! remove decayed particles CALL LUSPHE(SPH,APL) ! linear sphericity analysis IF(SPH.LT.0.) GOTO 100 ! too few particles in event for ! LUSPHE to work on it (unusual) CALL LUEDIT(31) ! orient event along axes above IF(IEVENT.EQ.1) CALL LULIST(2) ! list first treated event ..... ! fill analysis statistics CALL LUTHRU(THR,OBL) ! now do thrust analysis ..... ! more analysis statistics 100 CONTINUE ! CALL LUTABU(12) ! print particle composition ! statistics ..... ! print analysis statistics END \end{verbatim} \clearpage \section{Process Generation in PYTHIA} \label{s:PYTprocgen} Much can be said about the hard processes in {\Py} and the way they are generated. Therefore the material has been split into three sections. In the current one the philo\-sophy underlying the event generation scheme is presented. Here we provide a generic description, where some special cases are swept under the carpet. In the next section, the existing processes are enumerated, with some comments about applications and limitations. Finally, in the third section the generation routines and common block switches are described. The section starts with a survey of parton distributions, followed by a detailed description of the simple $2 \to 2$ and $2 \to 1$ hard subprocess generation schemes, including pairs of resonances. This is followed by a few comments on more complicated configurations. \subsection{Structure Functions} \label{ss:structfun} The parton-distribution function $f_i^a(x,Q^2)$ parametrizes the probability to find a parton $i$ with a fraction $x$ of the beam energy when the beam particle $a$ is probed by a hard scattering at virtuality scale $Q^2$. Usually the momentum-weighted combination $x f_i^a(x,Q^2)$ is used, for which the normalization condition $\sum_i \int_0^1 dx \, x f_i^a(x,Q^2) \equiv 1$ normally applies. The $Q^2$ dependence of parton distributions is perturbatively calculable, see section \ref{sss:initshowstruc}. The parton distributions in {\Py} come in many shapes, as shown in the following. \subsubsection{Baryons} For protons, many sets exist on the market. These are obtained by fits to experimental data, constrained so that the $Q^2$ dependence is in accordance with the standard QCD evolution equations. The (new) default in {\sc Pythia} is CTEQ2L \cite{Bot93}, a modern leading-order fit. Nine other sets are found in {\sc Pythia}. The complete list is: \begin{Itemize} \item EHLQ sets 1 and 2 \cite{Eic84}. \item DO sets 1 and 2 \cite{Duk82}. \item the other CTEQ2 fits, namely CTEQ2M, CTEQ2MS, CTEQ2MF, CTEQ2ML, and CTEQ2D \cite{Bot93}. \end{Itemize} Of these, EHLQ, DO and CTEQ2L are leading-order parton distributions, while CTEQ2D are in the next-to-leading-order DIS scheme and the rest in the next-to-leading order $\br{\mrm{MS}}$ scheme. The EHLQ and DO sets are by now rather old, and are kept mainly for backwards compatibility. Since only Born-level matrix elements are included in the program, there is no particular reason to use higher-order parton distributions --- the resulting combination is anyway only good to leading-order accuracy. (Some higher-order corrections are effectively included by the parton-shower treatment, but there is no exact match.) There is a steady flow of new parton-distribution sets on the market. To keep track of all of them is a major work on its own. Therefore {\Py} contains an interface to an external library of parton distribution functions, {\sc Pdflib} \cite{Plo93}. This is a truly encyclopedic collection of almost all proton, pion and photon parton distributions proposed since the late 70's. Two dummy routines come with the {\Py} package, so as to avoid problems with unresolved external references if {\sc Pdflib} is not linked. One should also note that {\Py} does not check the results, but assumes that sensible answers will be returned, also outside the nominal $(x, Q^2)$ range of a set. Only the sets that come with {\Py} have been suitably modified to provide reasonable answers outside their nominal domain of validity. From the proton parton distributions, those of the neutron are obtained by isospin conjugation, i.e. $f_{\u}^{\n} = f_{\d}^{\p}$ and $f_{\d}^{\n} = f_{\u}^{\p}$. The program does allow for incoming beams of a number of hyperons: $\Lambda^0$, $\Sigma^{-,0,+}$, $\Xi^{-,0}$ and $\Omega^-$. Here one has essentially no experimental information. One could imagine to construct models in which valence $\s$ quarks are found at larger average $x$ values than valence $\u$ and $\d$ ones, because of the larger $\s$-quark mass. However, hyperon beams is a little-used part of the program, included only for a few specific studies. Therefore a simple approach has been taken, in which an average valence quark distribution is constructed as $f_{\mrm{val}} = (f_{\u,\mrm{val}}^{\p} + f_{\d,\mrm{val}}^{\p})/3$, according to which each valence quark in a hyperon is assumed to be distributed. Sea-quark and gluon distributions are taken as in the proton. Any proton parton distribution set may be used with this procedure. \subsubsection{Mesons and photons} Data on meson parton distributions are scarce, so only very few sets have been constructed, and only for the $\pi^{\pm}$. {\Py} contains the Owens set 1 and 2 parton distributions \cite{Owe84}, which for a long time were essentially the only sets on the market. The former is the default in {\Py}. Recently more sets have appeared. These are found in {\sc Pdflib} and can therefore be used by {\Py}, just as described above for protons. Sets of photon parton distributions have been obtained as for hadrons; an additional complication comes from the necessity to handle the matching of the vector meson dominance (VMD) and the perturbative pieces in a consistent manner. {\Py} contains the Drees--Grassie set of parton distributions \cite{Dre85}, which therefore is default. In addition there is an interface to the {\sc Pdflib} library \cite{Plo93}, as for the proton. The sets above do not allow a subdivision of the photon parton distributions into one VMD part (from nonperturbative $\gamma \to V$ transitions, where $V$ is a $\rho^0$, $\omega$, $\phi$ or $\Jpsi$ vector meson) and one anomalous part (from perturbative $\gamma \to \q\qbar$ branchings). This subdivision is necessary for the more sophisticated photoproduction modelling in \cite{Sch93,Sch93a}, see section \ref{sss:photoprod}. Therefore, for the VMD part alone, the $\rho^0$ parton distribution is found from the assumed equality \begin{equation} f^{\rho^0}_i = f^{\pi^0}_i = \frac{1}{2} \, (f^{\pi^+}_i + f^{\pi^-}_i) ~. \end{equation} Thus any $\pi^+$ parton distribution set, from any library, can be turned into a VMD $\rho^0$ set. The $\omega$ parton distribution is assumed the same, while the $\phi$ and $\Jpsi$ ones are handled in the very crude approximation $f^{\phi}_{\s,\mrm{val}} = f^{\pi^+}_{\u,\mrm{val}}$ and $f^{\phi}_{\mrm{sea}} = f^{\pi^+}_{\mrm{sea}}$. The anomalous part of the photon parton distribution is entirely perturbatively calculable, given a lower cut-off scale. The leading order parametrization of \cite{Sch93c} is attached to the {\Py} program. There are no alternatives offered. \subsubsection{Leptons} \label{sss:estructfun} Contrary to the hadron case, there is no necessity to introduce the parton-distribution function concept for leptons. A lepton can be considered as a point-like particle, with initial-state radiation handled by higher-order matrix elements. However, the parton distribution function approach offers a slightly simplified but very economical description of initial-state radiation effects for any hard process, also those for which higher-order corrections are not yet calculated. Structure functions for electrons have been introduced in {\Py}, but not yet for muons, i.e. currently $f_{\mu}^{\mu}(x, Q^2) = \delta (x-1)$. Also for the electron one is free to use a simple `unresolved' $\e$, $f_{\e}^{\e}(x, Q^2) = \delta(x-1)$, where the $\e$ retains the full original momentum. Electron parton distributions are calculable entirely from first principles, but different levels of approximation may be used. The parton-distribution formulae in {\Py} are based on a next-to-leading-order exponentiated description, see ref. \cite{Kle89}, p. 34. The approximate behaviour is \begin{eqnarray} & & f_{\e}^{\e}(x,Q^2) \approx \frac{\beta}{2} (1-x)^{\beta/2-1} ~; \nonumber \\ & & \beta = \frac{2 \alphaem}{\pi} \left( \ln \frac{Q^2}{m_{\e}^2} -1 \right) ~. \end{eqnarray} The form is divergent but integrable for $x \to 1$, i.e. the electron likes to keep most of the energy. To handle the numerical precision problems for $x$ very close to unity, the parton distribution is set, by hand, to zero for $x > 0.999999$, and is rescaled upwards in the range $0.9999 < x < 0.999999$, in such a way that the total area under the parton distribution is preserved: \begin{equation} \left( f_{\e}^{\e}(x,Q^2) \right)_{\mrm{mod}} = \left\{ \begin{array}{ll} f_{\e}^{\e}(x,Q^2) & 0 \leq x \leq 0.9999 \\[2mm] \frac{\displaystyle 100^{\beta/2}}{\displaystyle 100^{\beta/2}-1} \, f_{\e}^{\e}(x,Q^2) & 0.9999 < x \leq 0.999999 \\[4mm] 0 & x > 0.999999 \, ~. \end{array} \right. \end{equation} The branchings $\e \to \e \gamma$, which are responsible for the softening of the $f_{\e}^{\e}$ parton distribution, also gives rise to a flow of photons. In photon-induced hard processes, the $f_{\gamma}^{\e}$ parton distribution can be used to describe the equivalent flow of photons. The formula used in the program is the simple first-order expression. There is some ambiguity in the choice of $Q^2$ range over which emissions should be included. The na\"{\i}ve (default) choice is \begin{equation} f_{\gamma}^{\e}(x,Q^2) = \frac{\alphaem}{2 \pi} \, \frac{1+(1-x)^2}{x} \, \ln \left( \frac{Q^2}{m_{\e}^2} \right) ~. \end{equation} Here it is assumed that only one scale enters the problem, namely that of the hard interaction, and that the scale of the branching $\e \to \e \gamma$ is bounded from above by the hard interaction scale. For a pure QCD or pure QED shower this is an appropriate procedure, cf. section \ref{sss:showermatching}, but in other cases it may not be optimal. In particular, for photoproduction the alternative that is probably most appropriate is \cite{Ali88}: \begin{equation} f_{\gamma}^{\e}(x,Q^2) = \frac{\alphaem}{2 \pi} \, \frac{1+(1-x)^2}{x} \, \ln \left( \frac{Q^2_{\mmax} (1-x)}{m_{\e}^2 \, x^2} \right) ~. \end{equation} Here $Q^2_{\mmax}$ is a user-defined cut for the range of scattered electron kinematics that is counted as photoproduction. Note that we now deal with two different $Q^2$ scales, one related to the hard subprocess itself, which appears as the argument of the parton distribution, and the other related to the scattering of the electron, which is reflected in $Q^2_{\mmax}$. In resolved photoproduction or resolved $\gamma\gamma$ interactions, one has to include the parton distributions for quarks and gluons inside the photon inside the electron. There are no published sets where results are directly presented in terms of quark and gluon distributions inside the electron. In the program, the $f_{\q,\g}^{\e}$ are therefore obtained by a numerical convolution according to \begin{equation} f_{\q,\g}^{\e}(x, Q^2) = \int_x^1 \frac{dx_{\gamma}}{x_{\gamma}} \, f_{\gamma}^{\e}(x_{\gamma}, Q^2) \, f^{\gamma}_{\q,\g} \! \left( \frac{x}{x_{\gamma}}, Q^2 \right) ~, \label{pg:foldqgine} \end{equation} with $f^{\e}_{\gamma}$ as discussed above. The necessity for numerical convolution makes this parton distribution evaluation rather slow compared with the others; one should therefore only have it switched on for resolved photoproduction studies. One can obtain the positron distribution inside an electron, which is also the electron sea parton distribution, by a convolution of the two branchings $\e \to \e \gamma$ and $\gamma \to \ee$; the result is \cite{Che75} \begin{equation} f_{\e^+}^{\e^-}(x,Q^2) = \frac{1}{2} \, \left\{ \frac{\alphaem}{2 \pi} \, \left( \ln \frac{Q^2}{m_{\e}^2} -1 \right) \right\}^2 \, \frac{1}{x} \, \left( \frac{4}{3} - x^2 - \frac{4}{3} x^3 + 2x(1+x) \ln x \right) ~. \end{equation} Finally, the program also contains the distribution of a transverse $\W^-$ inside an electron \begin{equation} f_{\W}^{\e}(x,Q^2) = \frac{\alphaem}{2 \pi} \, \frac{1}{4 \ssintw} \, \frac{1+(1-x)^2}{x} \, \ln \left( 1 + \frac{Q^2}{m_{\W}^2} \right) ~. \end{equation} \subsection{Kinematics and Cross section for a $2 \to 2$ Process} \label{ss:kinemtwo} In this section we begin the description of kinematics selection and cross-section calculation. The example is for the case of a $2 \to 2$ process, with final-state masses assumed to be vanishing. Later on we will expand to finite fixed masses, and to resonances. Consider two incoming beam particles in their c.m. frame, each with energy $E_{\mrm{beam}}$. The total squared c.m. energy is then $s = 4 E_{\mrm{beam}}^2$. The two partons that enter the hard interaction do not carry the total beam momentum, but only fractions $x_1$ and $x_2$, respectively, i.e. they have four-momenta \begin{eqnarray} p_1 & = & E_{\mrm{beam}}(x_1; 0, 0, x_1) ~, \nonumber \\ p_2 & = & E_{\mrm{beam}}(x_2; 0, 0, -x_2) ~. \end{eqnarray} There is no reason to put the incoming partons on the mass shell, i.e. to have time-like incoming four-vectors, since partons inside a particle are always virtual and thus space-like. These space-like virtualities are introduced as part of the initial-state parton-shower description, see section \ref{sss:initshowtrans}, but do not affect the formalism of this section. The one example where it would be appropriate to put a parton on the mass shell is for an incoming lepton beam, but even here the massless kinematics description is adequate as long as the c.m. energy is correctly calculated with masses. The squared invariant mass of the two partons is defined as \begin{equation} \hat{s} = (p_1 + p_2)^2 = x_1 \, x_2 \, s ~. \end{equation} Instead of $x_1$ and $x_2$, it is often customary to use $\tau$ and either $y$ or $x_F$: \begin{eqnarray} \tau & = & x_1 x_2 = \frac{\hat{s}}{s} ~; \\ y & = & \frac{1}{2} \ln \frac{x_1}{x_2} ~; \\ x_F & = & x_1 - x_2 ~. \end{eqnarray} In addition to $x_1$ and $x_2$, two additional variables are needed to describe the kinematics of a scattering $1 + 2 \to 3 + 4$. One corresponds to the azimuthal angle $\varphi$ of the scattering plane around the beam axis. This angle is always isotropically distributed for unpolarized incoming beam particles, and so need not be considered further. The other variable can be picked as $\hat{\theta}$, the polar angle of parton 3 in the c.m. frame of the hard scattering. The conventional choice is to use the variable \begin{equation} \hat{t} = (p_1 - p_3)^2 = (p_2 - p_4)^2 = - \frac{\hat{s}}{2} (1 - \cos \hat{\theta}) ~, \end{equation} with $\hat{\theta}$ defined as above. In the following, we will make use of both $\hat{t}$ and $\hat{\theta}$. It is also customary to define $\hat{u}$, \begin{equation} \hat{u} = (p_1 - p_4)^2 = (p_2 - p_3)^2 = - \frac{\hat{s}}{2} (1 + \cos \hat{\theta}) ~, \end{equation} but $\hat{u}$ is not an independent variable since \begin{equation} \hat{s} + \hat{t} + \hat{u} = 0 ~. \end{equation} If the two outgoing particles have masses $m_3$ and $m_4$, respectively, then the four-momenta in the c.m. frame of the hard interaction are given by \begin{equation} \hat{p}_{3,4} = \left( \frac{\hat{s} \pm (m_3^2-m_4^2)}{2\sqrt{\hat{s}}} , \pm \frac{\sqrt{\hat{s}}}{2} \, \beta_{34} \sin \hat{\theta}, 0, \pm \frac{\sqrt{\hat{s}}}{2} \, \beta_{34} \cos \hat{\theta} \right) ~, \end{equation} where \begin{equation} \beta_{34} = \sqrt{ \left( 1 - \frac{m_3^2}{\hat{s}} - \frac{m_4^2}{\hat{s}} \right)^2 - 4 \, \frac{m_3^2}{\hat{s}} \, \frac{m_4^2}{\hat{s}} } ~. \end{equation} Then $\hat{t}$ and $\hat{u}$ are modified to \begin{equation} \hat{t}, \hat{u} = - \frac{1}{2} \left\{ ( \hat{s} - m_3^2 - m_4^2 ) \mp \hat{s} \, \beta_{34} \cos \hat{\theta} \right\} ~, \end{equation} with \begin{equation} \hat{s} + \hat{t} + \hat{u} = m_3^2 + m_4^2 ~. \end{equation} The cross section for the process $1 + 2 \to 3 + 4$ may be written as \begin{eqnarray} \sigma & = & \int \int \int \d x_1 \, \d x_2 \, \d \hat{t} \, f_1(x_1, Q^2) \, f_2(x_2, Q^2) \, \frac{\d \hat{\sigma}}{\d \hat{t}} \nonumber \\ & = & \int \int \int \frac{\d \tau}{\tau} \, \d y \, \d \hat{t} \, x_1 f_1(x_1, Q^2) \, x_2 f_2(x_2, Q^2) \, \frac{\d \hat{\sigma}}{\d \hat{t}} ~. \label{pg:sigma} \end{eqnarray} The choice of $Q^2$ scale is ambiguous, and several alternatives are available in the program. For massless outgoing particles the default is the squared transverse momentum \begin{equation} Q^2 = \hat{p}_{\perp}^2 = \frac{\hat{s}}{4} \sin^2\hat{\theta} = \frac{\hat{t}\hat{u}}{\hat{s}} ~, \end{equation} which is modified to \begin{equation} Q^2 = \frac{1}{2} (m_{\perp 3}^2 + m_{\perp 4}^2) = \frac{1}{2} (m_3^2 + m_4^2) + \hat{p}_{\perp}^2 = \frac{1}{2} (m_3^2 + m_4^2) + \frac{\hat{t} \hat{u} - m_3^2 m_4^2}{\hat{s}} \end{equation} when masses are introduced. The mass term is selected such that, for $m_3 = m_4 = m$, the expression reduces to the squared transverse mass, $Q^2 = \hat{m}_{\perp}^2 = m^2 + \hat{p}_{\perp}^2$. The $\d \hat{\sigma}/\d \hat{t}$ expresses the differential cross section for a scattering, as a function of the kinematical quantities $\hat{s}$, $\hat{t}$ and $\hat{u}$. It is in this function that the physics of a given process resides. The performance of a machine is measured in terms of its luminosity $\cal L$, which is directly proportional to the number of particles in each bunch and to the bunch crossing frequency, and inversely proportional to the area of the bunches at the collision point. For a process with a $\sigma$ as given by eq. (\ref{pg:sigma}), the differential event rate is given by $\sigma {\cal L}$, and the number of events collected over a given period of time \begin{equation} N = \sigma \, \int {\cal L} \, \d t ~. \end{equation} The program does not calculate the number of events, but only the integrated cross sections. \subsection{Resonance Production} \label{ss:kinemreson} The simplest way to produce a resonance is by a $2 \to 1$ process. If the decay of the resonance is not considered, the cross-section formula does not depend on $\hat{t}$, but takes the form \begin{equation} \sigma = \int \int \frac{\d \tau}{\tau} \, \d y \, x_1 f_1(x_1, Q^2) \, x_2 f_2(x_2, Q^2) \, \hat{\sigma}(\hat{s}) ~. \label{pg:sigmares} \end{equation} Here the physics is contained in the cross section $\hat{\sigma}(\hat{s})$. The $Q^2$ scale is usually taken to be $Q^2 = \hat{s}$. In published formulae, cross sections are often given in the zero-width approximation, i.e. $\hat{\sigma}(\hat{s}) \propto \delta (\hat{s} - m_R^2)$, where $m_R$ is the mass of the resonance. Introducing the scaled mass $\tau_R = m_R^2/s$, this corresponds to a delta function $\delta (\tau - \tau_R)$, which can be used to eliminate the integral over $\tau$. However, what we normally want to do is replace the $\delta$ function by the appropriate Breit--Wigner shape. For a resonance width $\Gamma_R$ this is achieved by the replacement \begin{equation} \delta (\tau - \tau_R) \to \frac{s}{\pi} \, \frac{m_R \Gamma_R}{(s \tau - m_R^2)^2 + m_R^2 \Gamma_R^2} ~. \label{pg:resshapeone} \end{equation} In this formula the resonance width $\Gamma_R$ is a constant. An improved description of resonance shapes is obtained if the width is made $\hat{s}$-dependent (occasionally also referred to as mass-dependent width, since $\hat{s}$ is not always the resonance mass), see e.g. \cite{Ber89}. To first approximation, this means that the expression $m_R \Gamma_R$ is to be replaced by $\hat{s} \Gamma_R / m_R$. To be more precise, in the program the quantity $H_R(\hat{s})$ is introduced, and the Breit--Wigner is written as \begin{equation} \delta (\tau - \tau_R) \to \frac{s}{\pi} \, \frac{H_R(s \tau)}{(s \tau - m_R^2)^2 + H_R^2(s \tau)} ~. \label{pg:resshapetwo} \end{equation} The $H_R$ factor is evaluated as a sum over all possible final-state channels, $H_R = \sum_f H_R^{(f)}$. Each decay channel may have its own $\hat{s}$ dependence, as follows. A decay to a fermion pair, $R \to \f \fbar$, gives no contribution below threshold, i.e. for $\hat{s} < 4 m_{\f}^2$. Above threshold, $H_R^{(f)}$ is proportional to $\hat{s}$, multiplied by a threshold factor $\beta (3 - \beta^2)/2$ for the vector part of a spin 1 resonance, by $\beta^3$ for the axial vector part, and again by $\beta^3$ for a spin 0 resonance. Here $\beta = \sqrt{1 - 4m_{\f}^2/\hat{s}}$. For the decay into unequal masses, e.g. of the $\W^+$, corresponding but more complicated expressions are used. For decays into a quark pair, the universal first-order strong correction factor $1 + \alphas(\hat{s}) / \pi$ is included in $H_R^{(f)}$. The second-order corrections are often known, but then are specific to each resonance, and are not included. An option exists for the $\gamma/\Z^0/\Z'^0$ resonances, where threshold effects due to $\q\qbar$ bound-state formation are taken into account in a smeared-out, average sense, see eq. (\ref{pp:threshenh}). For other decay channels, not into fermion pairs, the $\hat{s}$ dependence is typically more complicated. For instance, the decay $\H^0 \to \W^+ \W^-$ has a partial width proportional to $\hat{s}^2$, with a threshold factor $\beta^3$. Since a Higgs with $m_{\H} < 2 m_{\W}$ could still decay in this channel, it is in fact necessary to perform a two-dimensional integral over the $W^{\pm}$ Breit--Wigner mass distributions to obtain the correct result (and this has to be done numerically, at least in part). Fortunately, a Higgs particle lighter than $2 m_{\W}$ is sufficiently narrow that the integral only needs to be performed once and for all at initialization (whereas most other partial widths are recalculated whenever needed). Channels that proceed via loops, such as $\H \to \g \g$, also display complicated threshold behaviours. The coupling structure within the electroweak sector is usually (re)expressed in terms of gauge boson masses, $\alphaem$ and $\ssintw$, i.e. factors of $G_F$ are replaced according to \begin{equation} \sqrt{2} G_F = \frac{\pi \, \alphaem}{\ssintw \, m_{\W}^2} ~. \end{equation} Having done that, $\alphaem$ is allowed to run \cite{Kle89}, and is evaluated at the $\hat{s}$ scale. Thereby the relevant electroweak loop correction factors are recovered at the $m_{\W}/m_{\Z}$ scale. Currently $\ssintw$ is not allowed to run. For the Higgs particle, the couplings to fermions are proportional to the fermion masses; then also the masses are evaluated at the $\hat{s}$ scale. In summary, we see that an $\hat{s}$ dependence may enter several different ways into the $H_R^{(f)}$ expressions from which the total $H_R$ is built up. Also note that, with the exception of the term $(\s \tau - m_R^2)^2$ in the denominator of the Breit--Wigner, no memory remains of the nominal $m_R$ mass: everywhere else, what enters is the actual resonance mass $\sqrt{\hat{s}}$. When only decays to a specific final state $f$ are considered, the $H_R$ in the denominator remains the sum over all allowed decay channels, but the numerator only contains the $H_R^{(f)}$ term of the final state considered. If the combined production and decay process $i \to R \to f$ is considered, the same $\hat{s}$ dependence is implicit in the coupling structure of $i \to R$ as one would have had in $R \to i$, i.e. to first approximation there is a symmetry between couplings of a resonance to the initial and to the final state. The cross section $\hat{\sigma}$ is therefore, in the program, written in the form \begin{equation} \hat{\sigma}_{i \to R \to f}(\hat{s}) \propto \frac{\pi}{\hat{s}} \, \frac{H_R^{(i)}(\hat{s}) \, H_R^{(f)}(\hat{s})} {(\hat{s} - m_R^2)^2 + H_R^2(\hat{s})} ~. \label{pg:Hinoutsym} \end{equation} As a simple example, the cross section for the process $\e^- \br{\nu}_{\e} \to \W^- \to \mu^- \br{\nu}_{\mu}$ can be written as \begin{equation} \hat{\sigma}(\hat{s}) = 12 \, \frac{\pi}{\hat{s}} \, \frac{H_{\W}^{(i)}(\hat{s}) \, H_{\W}^{(f)}(\hat{s})} {(\hat{s} - m_{\W}^2)^2 + H_{\W}^2(\hat{s})} ~, \end{equation} where \begin{equation} H_{\W}^{(i)}(\hat{s}) = H_{\W}^{(f)}(\hat{s}) = \frac{\alphaem(\hat{s})}{24 \, \ssintw} \, \hat{s} ~. \end{equation} If the effects of several initial and/or final states are studied, it is straightforward to introduce an appropriate summation in the numerator. The analogy between the $H_R^{(f)}$ and $H_R^{(i)}$ cannot be pushed too far, however. The two differ in several important aspects. Firstly, colour factors appear reversed: the decay $R \to \q\qbar$ contains a colour factor $N_C = 3$ enhancement, while $\q\qbar \to R$ is instead suppressed by a factor $1/N_C = 1/3$. Secondly, the $1 + \alphas(\hat{s}) / \pi$ first-order correction factor for the final state has to be replaced by a more complicated $K$ factor for the initial state. This factor is not usually known, or it is known (to first non-trivial order) but too lengthy to be included in the program. Thirdly, incoming partons as a rule are space-like. All the threshold suppression factors of the final state expressions are therefore irrelevant when production is considered. In sum, the $H_R^{(f)}$--$H_R^{(i)}$ analogy is mainly useful as a consistency cross-check, while the two usually are calculated separately. Exceptions include the rather messy loop structure involved in $\g\g \to \H^0$ and $\H^0 \to \g\g$, which is only coded once. It is of some interest to consider the observable resonance shape when the effects of parton distributions are included. In a hadron collider, to first approximation, parton distributions tend to have a behaviour roughly like $f(x) \propto 1/x$ for small $x$ --- this is why $f(x)$ is replaced by $xf(x)$ in eq. (\ref{pg:sigma}). Instead, the basic parton-distribution behaviour is shifted into the factor of $1/\tau$ in the integration phase space $\d \tau/\tau$, cf. eq. (\ref{pg:sigmares}). When folded with the Breit--Wigner shape, two effects appear. One is that the overall resonance is tilted: the low-mass tail is enhanced and the high-mass one suppressed. The other is that an extremely long tail develops on the low-mass side of the resonance: when $\tau \to 0$, eq. (\ref{pg:Hinoutsym}) with $H_R(\hat{s}) \propto \hat{s}$ gives a $\hat{\sigma}(\hat{s}) \propto \hat{s} \propto \tau$, which exactly cancels the $1/\tau$ factor mentioned above. Na\"{\i}vely, the integral over $y$, $\int \d y = - \ln \tau$, therefore gives a net logarithmic divergence of the resonance shape when $\tau \to 0$. Clearly, it is then necessary to consider the shape of the parton distributions in more detail. At not-too-small $Q^2$, the evolution equations in fact lead to parton distributions more strongly peaked than $1/x$, typically with $xf(x) \propto x^{-0.3}$, and therefore a divergence like $\tau^{-0.3}$ in the cross-section expression. Eventually this divergence is regularized by a closing of the phase space, i.e. that $H_R(\hat{s})$ vanishes faster than $\hat{s}$, and by a less drastic small-$x$ parton-distribution behaviour when $Q^2 \approx \hat{s} \to 0$. The secondary peak at small $\tau$ may give a rather high cross section, which can even rival that of the ordinary peak around the nominal mass. This is the case, for instance, with $\W$ production. Such a peak has never been observed experimentally, but this is not surprising, since the background from other processes is overwhelming at low $\hat{s}$. Thus a lepton of one or a few GeV of transverse momentum is far more likely to come from the decay of a charm or bottom hadron than from a `$\W$' of a mass of a few GeV. When resonance production is studied, it is therefore important to set limits on the mass of the resonance, so as to agree with the experimental definition, at least to first approximation. If not, cross-section information given by the program may be very confusing. In $\ee$ colliders, the $f_{\e}^{\e}$ parton distribution is peaked at $x = 1$ rather than at $x = 0$. The situation therefore is the opposite, if one considers e.g. $\Z^0$ production in a machine running at energies above $m_{\Z}$: the tail towards lower masses is suppressed and the one towards higher masses enhanced, with a sharp secondary peak at around the nominal energy of the machine. Also in this case, an appropriate definition of cross sections is therefore necessary --- with additional complications due to the interference between $\gamma^*$ and $\Z^0$. In $2 \to 2$ processes where a pair of resonances are produced, e.g. $\ee \to \Z^0 \H^0$, cross section are almost always given in the zero-width approximation for the resonances. Here two substitutions of the type \begin{equation} 1 = \int \delta (m^2 - m_R^2) \, dm^2 \to \int \frac{1}{\pi} \, \frac{m_R \Gamma_R}{(m^2 - m_R^2)^2 + m_R^2 \Gamma_R^2} \, dm^2 \label{pg:resshapethree} \end{equation} are used to introduce mass distributions for the two resonance masses, i.e. $m_3^2$ and $m_4^2$. In the formula, $m_R$ is the nominal mass and $m$ the actually selected one. The phase-space integral over $x_1$, $x_1$ and $\hat{t}$ in eq. (\ref{pg:sigma}) is then extended to involve also $m_3^2$ and $m_4^2$. The effects of the mass-dependent width is only partly taken into account, by replacing the nominal masses $m_3^2$ and $m_4^2$ in the $\d \hat{\sigma} / \d \hat{t}$ expression by the actually generated ones (also e.g. in the relation between $\hat{t}$ and $\cos\hat{\theta}$), while the widths are evaluated at the nominal masses. This is the equivalent of a simple replacement of $m_R \Gamma_R$ by $\hat{s} \Gamma_R / m_R$ in the numerator of eq. (\ref{pg:resshapeone}), but not in the denominator. In addition, the full threshold dependence, i.e. the $\beta$-dependent factors, is not reproduced. There is no particular reason why the full mass-dependence could not be introduced, except for the extra work and time consumption needed for each process. In fact, the matrix elements for several $\gammaZ$ production processes do contain the full expressions. On the other hand, the matrix elements given in the literature are often valid only when the resonances are almost on the mass shell, since some graphs have been omitted. As an example, the process $\q\qbar \to \e^- \br{\nu}_{\e} \mu^+ \nu_{\mu}$ is dominated by $\q\qbar \to \W^- \W^+$ when each of the two lepton pairs is close to $m_{\W}$ in mass, but in general also receives contributions e.g. from $\q\qbar \to \Z^0 \to \ee$, followed by $\e^+ \to \br{\nu}_{\e} \W^+$ and $\W^+ \to \mu^+ \nu_{\mu}$. The latter contributions are neglected in cross sections given in the zero-width approximation. Processes with one final-state resonance and another ordinary final-state product, e.g. $\q \g \to \W^+ \q'$, are treated in the same spirit as the $2 \to 2$ processes with two resonances, except that only one mass need be selected according to a Breit--Wigner. \subsection{Cross-section Calculations} \label{ss:PYTcrosscalc} In the program, the variables used in the generation of a $2 \to 2$ process are $\tau$, $y$ and $z = \cos\hat{\theta}$. For a $2 \to 1$ process, the $z$ variable can be integrated out, and need therefore not be generated as part of the hard process, except when the allowed angular range of decays is restricted. In unresolved lepton beams, i.e. when $f_{\e}^{\e}(x) = \delta(x-1)$, the variables $\tau$ and/or $y$ may be integrated out. We will cover all these special cases towards the end of the section, and here concentrate on `standard' $2 \to 2$ and $2 \to 1$ processes. \subsubsection{The simple $2 \to 2$ processes} In the spirit of section \ref{ss:MCdistsel}, we want to select simple functions such that the true $\tau$, $y$ and $z$ dependence of the cross sections is approximately modelled. In particular, (almost) all conceivable kinematical peaks should be represented by separate terms in the approximate formulae. If this can be achieved, the ratio of the correct to the approximate cross sections will not fluctuate too much, but allow reasonable Monte Carlo efficiency. Therefore the variables are generated according to the distributions $h_{\tau}(\tau)$, $h_y(y)$ and $h_z(z)$, where normally \begin{eqnarray} h_{\tau}(\tau) & = & \frac{c_1}{{\cal I}_1} \, \frac{1}{\tau} + \frac{c_2}{{\cal I}_2} \, \frac{1}{\tau^2} + \frac{c_3}{{\cal I}_3} \, \frac{1}{\tau(\tau + \tau_R)} + \frac{c_4}{{\cal I}_4} \, \frac{1}{(s\tau - m_R^2)^2 + m_R^2 \Gamma_R^2} \nonumber \\ & & + \frac{c_5}{{\cal I}_5} \, \frac{1}{\tau(\tau + \tau_{R'})} + \frac{c_6}{{\cal I}_6} \, \frac{1}{(s\tau - m_{R'}^2)^2 + m_{R'}^2 \Gamma_{R'}^2} ~, \\[1mm] h_y(y) & = & \frac{c_1}{{\cal I}_1} \, (y - y_{\mmin}) + \frac{c_2}{{\cal I}_2} \, (y_{\mmax} - y) + \frac{c_3}{{\cal I}_3} \, \frac{1}{\cosh y} ~, \\[1mm] h_z(z) & = & \frac{c_1}{{\cal I}_1} + \frac{c_2}{{\cal I}_2} \, \frac{1}{a-z} + \frac{c_3}{{\cal I}_3} \, \frac{1}{a+z} + \frac{c_4}{{\cal I}_4} \, \frac{1}{(a-z)^2} + \frac{c_5}{{\cal I}_5} \, \frac{1}{(a+z)^2} ~. \label{pg:hforms} \end{eqnarray} Here each term is separately integrable, with an invertible primitive function, such that generation of $\tau$, $y$ and $z$ separately is a standard task, as described in section \ref{ss:MCdistsel}. In the following we describe the details of the scheme, including the meaning of the coefficients $c_i$ and ${\cal I}_i$, which are separate for $\tau$, $y$ and $z$. The first variable to be selected is $\tau$. The range of allowed values, $\tau_{\mmin} \leq \tau \leq \tau_{\mmax}$, is generally constrained by a number of user-defined requirements. A cut on the allowed mass range is directly reflected in $\tau$, a cut on the $\pT$ range indirectly. The first two terms of $h_{\tau}$ are intended to represent a smooth $\tau$ dependence, as generally obtained in processes which do not receive contributions from $s$-channel resonances. Also $s$-channel exchange of essentially massless particles ($\gamma$, $\g$, light quarks and leptons) are accounted for, since these do not produce any separate peaks at non-vanishing $\tau$. The last four terms of $h_{\tau}$ are there to catch the peaks in the cross section from resonance production. These terms are only included when needed. Each resonance is represented by two pieces, a first to cover the interference with graphs which peak at $\tau =0$, plus the variation of parton distributions, and a second to approximate the Breit--Wigner shape of the resonance itself. The subscripts $R$ and $R'$ denote values pertaining to the two resonances, with $\tau_R = m_R^2/s$. Currently there is only one process where the full structure with two resonances is used, namely $\f \fbar \to \gamma^*/\Z^0/\Z'^0$. Otherwise either one or no resonance peak is taken into account. The kinematically allowed range of $y$ values is constrained by $\tau$, $|y| \leq - \frac{1}{2} \ln\tau$, and you may impose additional cuts. Therefore the allowed range $y_{\mmin} \leq y \leq y_{\mmax}$ is only constructed after $\tau$ has been selected. The first two terms of $h_y$ give a fairly flat $y$ dependence --- for processes which are symmetric in $y \leftrightarrow -y$, they will add to give a completely flat $y$ spectrum between the allowed limits. In principle, the natural subdivision would have been one term flat in $y$ and one forward--backward asymmetric, i.e. proportional to $y$. The latter is disallowed by the requirement of positivity, however. The $y - y_{\mmin}$ and $y_{\mmax} - y$ terms actually used give the same amount of freedom, but respect positivity. The third term is peaked at around $y = 0$, and represents the bias of parton distributions towards this region. The allowed $z = \cos\hat{\theta}$ range is na\"{\i}vely $-1 \leq z \leq 1$. However, most cross sections are divergent for $z \to \pm 1$, so some kind of regularization is necessary. Normally one requires $\pT \geq \pTmin$, which translates into $z^2 \leq 1 - 4 \pTmin^2/(\tau s)$ for massless outgoing particles. Since again the limits depend on $\tau$, the selection of $z$ is done after that of $\tau$. Additional requirements may constrain the range further. In particular, a $p_{\perp\mmax}$ constraint may split the allowed $z$ range into two, i.e. $z_{-\mmin} \leq z \leq z_{-\mmax}$ or $z_{+\mmin} \leq z \leq z_{+\mmax}$. An unsplit range is represented by $z_{-\mmax} = z_{+\mmin} = 0$. For massless outgoing particles the parameter $a = 1$ in $h_z$, such that the five terms represent a piece flat in angle and pieces peaked as $1/\hat{t}$, $1/\hat{u}$, $1/\hat{t}^2$, and $1/\hat{u}^2$, respectively. For non-vanishing masses one has $a = 1 + 2 m_3^2 m_4^2/\hat{s}^2$. In this case, the full range $-1 \leq z \leq 1$ is therefore available --- physically, the standard $\hat{t}$ and $\hat{u}$ singularities are regularized by the masses $m_3$ and $m_4$. For each of the terms, the ${\cal I}_i$ coefficients represent the integral over the quantity multiplying the coefficient $c_i$; thus, for instance: \begin{eqnarray} h_{\tau}: & & {\cal I}_1 = \int \frac{\d \tau}{\tau} = \ln \left( \frac{\tau_{\mmax}}{\tau_{\mmin}} \right) ~, \nonumber \\ & & {\cal I}_2 = \int \frac{\d \tau}{\tau^2} = \frac{1}{\tau_{\mmin}} - \frac{1}{\tau_{\mmax}} ~; \nonumber \\ h_y: & & {\cal I}_1 = \int (y - y_{\mmin}) \, \d y = \frac{1}{2} (y_{\mmax} - y_{\mmin})^2 ~; \nonumber \\ h_z: & & {\cal I}_1 = \int \d z = (z_{-\mmax} - z_{-\mmin}) + (z_{+\mmax} - z_{+\mmin}) , \nonumber \\ & & {\cal I}_2 = \int \frac{\d z}{a-z} = \ln \left( \frac{(a-z_{-\mmin})(a-z_{+\mmin})} {(a-z_{-\mmax})(a-z_{-\mmin})} \right) ~. \end{eqnarray} The $c_i$ coefficients are normalized to unit sum for $h_{\tau}$, $h_y$ and $h_z$ separately. They have a simple interpretation, as the probability for each of the terms to be used in the preliminary selection of $\tau$, $y$ and $z$, respectively. The variation of the cross section over the allowed phase space is explored in the initialization procedure of a {\Py} run, and based on this knowledge the $c_i$ are optimized so as to give functions $h_{\tau}$, $h_y$ and $h_z$ that closely follow the general behaviour of the true cross section. For instance, the coefficient $c_4$ in $h_{\tau}$ is to be made larger the more the total cross section is dominated by the region around the resonance mass. The phase-space points tested at initialization are put on a grid, with the number of points in each dimension given by the number of terms in the respective $h$ expression, and with the position of each point given by the median value of the distribution of one of the terms. For instance, the $\d \tau / \tau$ distribution gives a median point at $\sqrt{\tau_{\mmin}\tau_{\mmax}}$, and $\d \tau / \tau^2$ has the median $2 \tau_{\mmin} \tau_{\mmax} / (\tau_{\mmin} + \tau_{\mmax})$. Since the allowed $y$ and $z$ ranges depend on the $\tau$ value selected, then so do the median points defined for these two variables. With only a limited set of phase-space points studied at the initialization, the `optimal' set of coefficients is not uniquely defined. To be on the safe side, 40\% of the total weight is therefore assigned evenly between all allowed $c_i$, whereas the remaining 60\% are assigned according to the relative importance surmised, under the constraint that no coefficient is allowed to receive a negative contribution from this second piece. After a preliminary choice has been made of $\tau$, $y$ and $z$, it is necessary to find the weight of the event, which is to be used to determine whether to keep it or generate another one. Using the relation $\d \hat{t} = \hat{s} \, \beta_{34} \, \d z / 2$, eq. (\ref{pg:sigma}) may be rewritten as \begin{eqnarray} \sigma & = & \int \int \int \frac{\d \tau}{\tau} \, \d y \, \frac{\hat{s} \beta_{34}}{2} \d z \, x_1 f_1(x_1, Q^2) \, x_2 f_2(x_2, Q^2) \, \frac{\d \hat{\sigma}}{\d \hat{t}} \nonumber \\[1mm] & = & \frac{\pi}{s} \int h_{\tau}(\tau) \, \d \tau \int h_y(y) \, \d y \int h_z(z) \, \d z \, \beta_{34} \, \frac{x_1 f_1(x_1, Q^2) \, x_2 f_2(x_2, Q^2)} {\tau^2 h_{\tau}(\tau) \, h_y(y) \, 2 h_z(z)} \, \frac{\hat{s}^2}{\pi} \frac{\d \hat{\sigma}}{\d \hat{t}} \nonumber \\[1mm] & = & \left\langle \frac{\pi}{s} \, \frac{\beta_{34}}{\tau^2 h_{\tau}(\tau) \, h_y(y) \, 2 h_z(z)} \, x_1 f_1(x_1, Q^2) \, x_2 f_2(x_2, Q^2) \, \frac{\hat{s}^2}{\pi} \frac{\d \hat{\sigma}}{\d \hat{t}} \right\rangle ~. \label{pg:sigmamap} \end{eqnarray} In the middle line, a factor of $1 = h_{\tau}/h_{\tau}$ has been introduced to rewrite the $\tau$ integral in terms of a phase space of unit volume: $\int h_{\tau}(\tau) \, \d \tau = 1$ according to the relations above. Correspondingly for the $y$ and $z$ integrals. In addition, factors of $1 = \hat{s}/ (\tau s)$ and $1 = \pi / \pi$ are used to isolate the dimensionless cross section $(\hat{s}^2/\pi) \, \d \hat{\sigma} / \d \hat{t}$. The content of the last line is that, with $\tau$, $y$ and $z$ selected according to the expressions $h_{\tau}(\tau)$, $h_y(y)$ and $h_z(z)$, respectively, the cross section is obtained as the average of the final expression over all events. Since the $h$'s have been picked to give unit volume, there is no need to multiply by the total phase-space volume. As can be seen, the cross section for a given Monte Carlo event is given as the product of four factors, as follows: \begin{Enumerate} \item The $\pi/s$ factor, which is common to all events, gives the overall dimensions of the cross section, in GeV$^{-2}$. Since the final cross section is given in units of mb, the conversion factor of 1 GeV$^{-2} = 0.3894$ mb is also included here. \item Next comes the `Jacobian', which compensates for the change from the original to the final phase-space volume. \item The parton-distribution function weight is obtained by making use of the parton distribution libraries in {\Py} or externally. The $x_1$ and $x_2$ values are obtained from $\tau$ and $y$ via the relations $x_{1,2} = \sqrt{\tau} \exp(\pm y)$. \item Finally, the dimensionless cross section $(\hat{s}^2/\pi) \, \d \hat{\sigma} / \d \hat{t}$ is the quantity that has to be coded for each process separately, and where the physics content is found. \end{Enumerate} Of course, the expression in the last line is not strictly necessary to obtain the cross section by Monte Carlo integration. One could also have used eq. (\ref{pg:sigma}) directly, selecting phase-space points evenly in $\tau$, $y$ and $\hat{t}$, and averaging over those Monte Carlo weights. Clearly this would be much simpler, but the price to be paid is that the weights of individual events could fluctuate wildly. For instance, if the cross section contains a narrow resonance, the few phase-space points that are generated in the resonance region obtain large weights, while the rest do not. With our procedure, a resonance would be included in the $h_{\tau}(\tau)$ factor, so that more events would be generated at around the appropriate $\tau_R$ value (owing to the $h_{\tau}$ numerator in the phase-space expression), but with these events assigned a lower, more normal weight (owing to the factor $1/h_{\tau}$ in the weight expression). Since the weights fluctuate less, fewer phase-space points need be selected to get a reasonable cross-section estimate. In the program, the cross section is obtained as the average over all phase-space points generated. The events actually handed on to the user should have unit weight, however (an option with weighted events exists, but does not represent the mainstream usage). At initialization, after the $c_i$ coefficients have been determined, a search inside the allowed phase-space volume is therefore made to find the maximum of the weight expression in the last line of eq. (\ref{pg:sigmamap}). In the subsequent generation of events, a selected phase-space point is then retained with a probability equal to the weight in the point divided by the maximum weight. Only the retained phase-space points are considered further, and generated as complete events. The search for the maximum is begun by evaluating the weight in the same grid of points as used to determine the $c_i$ coefficients. The point with highest weight is used as starting point for a search towards the maximum. In unfortunate cases, the convergence could be towards a local maximum which is not the global one. To somewhat reduce this risk, also the grid point with second-highest weight is used for another search. After initialization, when events are generated, a warning message will be given by default at any time a phase-space point is selected where the weight is larger than the maximum, and thereafter the maximum weight is adjusted to reflect the new knowledge. This means that events generated before this time have a somewhat erroneous distribution in phase space, but if the maximum violation is rather modest the effects should be negligible. The estimation of the cross section is not affected by any of these considerations, since the maximum weight does not enter into eq. (\ref{pg:sigmamap}). For $2 \to 2$ processes with identical final-state particles, the symmetrization factor of $1/2$ is explicitly included at the end of the $\d \hat{\sigma} / \d \hat{t}$ calculation. In the final cross section, a factor of 2 is retrieved because of integration over the full phase space (rather than only half of it). That way, no special provisions are needed in the phase-space integration machinery. \subsubsection{Resonance production} We have now covered the simple $2 \to 2$ case. In a $2 \to 1$ process, the $\hat{t}$ integral is absent, and the differential cross section $\d \hat{\sigma} / \d \hat{t}$ is replaced by $\hat{\sigma}(\hat{s})$. The cross section may now be written as \begin{eqnarray} \sigma & = & \int \int \frac{\d \tau}{\tau} \, \d y \, x_1 f_1(x_1, Q^2) \, x_2 f_2(x_2, Q^2) \, \hat{\sigma}(\hat{s}) \nonumber \\ & = & \frac{\pi}{s} \int h_{\tau}(\tau) \, \d \tau \int h_y(y) \, \d y \, \frac{x_1 f_1(x_1, Q^2) \, x_2 f_2(x_2, Q^2)} {\tau^2 h_{\tau}(\tau) \, h_y(y)} \, \frac{\hat{s}}{\pi} \hat{\sigma}(\hat{s}) \nonumber \\ & = & \left\langle \frac{\pi}{s} \, \frac{1}{\tau^2 h_{\tau}(\tau) \, h_y(y)} \, x_1 f_1(x_1, Q^2) \, x_2 f_2(x_2, Q^2) \, \frac{\hat{s}}{\pi} \hat{\sigma}(\hat{s}) \right\rangle ~. \label{pg:sigmamapres} \end{eqnarray} The structure is thus exactly the same, but the $z$-related pieces are absent, and the r\^ole of the dimensionless cross section is played by $(\hat{s}/\pi) \hat{\sigma}(\hat{s})$. If the range of allowed decay angles of the resonance is restricted, e.g. by requiring the decay products to have a minimum transverse momentum, effectively this translates into constraints on the $z = \cos\hat{\theta}$ variable of the $2 \to 2$ process. The difference is that the angular dependence of a resonance decay is trivial, and that therefore the $z$-dependent factor can be easily evaluated. For a spin-0 resonance, which decays isotropically, the relevant weight is simply $(z_{-\mmax} - z_{-\mmin})/2 + (z_{+\mmax} - z_{+\mmin})/2$. For a transversely polarized spin-1 resonance the expression is, instead, \begin{equation} \frac{3}{8}(z_{-\mmax} - z_{-\mmin}) + \frac{3}{8}(z_{+\mmax} - z_{+\mmin}) + \frac{1}{8}(z_{-\mmax} - z_{-\mmin})^3 + \frac{1}{8}(z_{+\mmax} - z_{+\mmin})^3 ~. \end{equation} Since the allowed $z$ range could depend on $\tau$ and/or $y$ (it does for a $\pT$ cut), the factor has to be evaluated for each individual phase-space point and included in the expression of eq. (\ref{pg:sigmamapres}). For $2 \to 2$ processes where either of the final-state particles is a resonance, or both, an additional choice has to be made for each resonance mass, eq. (\ref{pg:resshapethree}). Since the allowed $\tau$, $y$ and $z$ ranges depend on $m_3^2$ and $m_4^2$, the selection of masses have to precede the choice of the other phase-space variables. Just as for the other variables, masses are not selected uniformly over the allowed range, but are rather distributed according to a function $h_m(m^2) \, dm^2$, with a compensating factor $1/h_m(m^2)$ in the `Jacobian'. The functional form picked is normally \begin{equation} h_m(m^2) = \frac{c_1}{{\cal I}_1} \, \frac{1}{\pi} \, \frac{m_R \Gamma_R}{(m^2 - m_R^2)^2 + m_R^2 \Gamma_R^2} + \frac{c_2}{{\cal I}_2} + \frac{c_3}{{\cal I}_3} \, \frac{1}{m^2} + \frac{c_4}{{\cal I}_4} \, \frac{1}{m^4} ~. \label{pg:reshdist} \end{equation} The definition of the ${\cal I}_i$ integrals is analogous to the one before. The $c_i$ coefficients are not found by optimization, but predetermined, normally to $c_1 = 0.8$, $c_2 = c_3 =0.1$, $c_4 = 0$. Clearly, had the phase space and the cross section been independent of $m_3^2$ and $m_4^2$, the optimal choice would have been to put $c_1 = 1$ and have all other $c_i$ vanishing --- then the $1/h_m$ factor of the `Jacobian' would exactly have cancelled the Breit--Wigner of eq. (\ref{pg:resshapethree}) in the cross section. The second and the third terms are there to cover the possibility that the cross section does not die away quite as fast as given by the na\"{\i}ve Breit--Wigner shape. In particular, the third term covers the possibility of a secondary peak at small $m^2$, in a spirit slightly similar to the one discussed for resonance production in $2 \to 1$ processes. The fourth term is only used for the process $\f \fbar \to (\gammaZ)(\gammaZ)$, where the $\gamma$ propagator guarantees that the cross section does have a significant secondary peak for $m^2 \to 0$. Therefore here the choice is $c_1 = 0.4$, $c_2 = 0.05$, $c_3 = 0.3$ and $c_4 = 0.25$. A few special tricks have been included to improve efficiency when the allowed mass range of resonances is constrained by kinematics or by user cuts. For instance, if a pair of equal or charge-conjugate resonances are produced, such as in $\ee \to \W^+ \W^-$, use is made of the constraint that the lighter of the two has to have a mass smaller than half the c.m. energy. \subsubsection{Lepton beams} Lepton beams have to be handled slightly differently from what has been described so far. One also has to distinguish between a lepton for which parton distributions are included and one which is treated as an unresolved point-like particle. The necessary modifications are the same for $2 \to 2$ and $2 \to 1$ processes, however, since the $\hat{t}$ degree of freedom is unaffected. If one incoming beam is an unresolved lepton, the corresponding parton-distribution piece collapses to a $\delta$ function. This function can be used to integrate out the $y$ variable: $\delta(x_{1,2} - 1) = \delta(y \pm (1/2) \ln \tau)$. It is therefore only necessary to select the $\tau$ and the $z$ variables according to the proper distributions, with compensating weight factors, and only one set of parton distributions has to be evaluated explicitly. If both incoming beams are unresolved leptons, both the $\tau$ and the $y$ variables are trivially given: $\tau = 1$ and $y = 0$. Structure function weights disappear completely. For a $2 \to 2$ process, only the $z$ selection remains to be performed, while a $2 \to 1$ process is completely specified, i.e. the cross section is a simple number that only depends on the c.m. energy. For a resolved electron, the $f_{\e}^{\e}$ parton distribution is strongly peaked towards $x = 1$. This affects both the $\tau$ and the $y$ distributions, which are not well described by either of the pieces in $h_{\tau}(\tau)$ or $h_y(y)$ in processes with interacting $\e^{\pm}$. (Processes which involve e.g. the $\gamma$ content of the $\e$ are still well simulated, since $f_{\gamma}^{\e}$ is peaked at small $x$.) If both parton distributions are peaked close to 1, the $h_{\tau}(\tau)$ expression in eq. (\ref{pg:hforms}) is therefore increased with one additional term of the form $h_{\tau}(\tau) \propto 1 / (1 - \tau)$, with coefficients $c_7$ and ${\cal I}_7$ determined as before. The divergence when $\tau \to 1$ is cut off by our regularization procedure for the $f_{\e}^{\e}$ parton distribution; therefore we only need consider $\tau < 1 - 2 \times 10^{-6}$. Correspondingly, the $h_y(y)$ expression is expanded with a term $1/(1 - \exp(y-y_0))$ when incoming beam number 1 consists of a resolved $\e^{\pm}$, and with a term $1/(1 - \exp(-y-y_0))$ when incoming beam number 2 consists of a resolved $\e^{\pm}$. Both terms are present for an $\ee$ collider, only one for an $\ep$ one. The coefficient $y_0 = - (1/2) \ln \tau$ is the na\"{\i}ve kinematical limit of the $y$ range, $|y| < y_0$. From the definitions of $y$ and $y_0$ it is easy to see that the two terms above correspond to $1/(1-x_1)$ and $1/(1-x_2)$, respectively, and thus are again regularized by our parton-distribution function cut-off. Therefore the integration ranges are $y < y_0 -10^{-6}$ for the first term and $y > - y_0 + 10^{-6}$ for the second one. \subsubsection{Mixing processes} \label{sss:mixingproc} In the cross-section formulae given so far, we have deliberately suppressed a summation over the allowed incoming flavours. For instance, the process $\f\fbar \to \Z^0$ in a hadron collider receives contributions from $\u\ubar \to \Z^0$, $\d\dbar \to \Z^0$, $\s\sbar \to \Z^0$, and so on. These contributions share the same basic form, but differ in the parton-distribution weights and (usually) in a few coupling constants in the hard matrix elements. It it therefore convenient to generate the terms together, as follows: \begin{Enumerate} \item A phase-space point is picked, and all common factors related to this choice are evaluated, i.e. the `Jacobian' and the common pieces of the matrix elements (e.g. for a $\Z^0$ the basic Breit--Wigner shape, excluding couplings to the initial flavour). \item The parton-distribution-function library is called to produce all the parton distributions, at the relevant $x$ and $Q^2$ values, for the two incoming beams. \item A loop is made over the two incoming flavours, one from each beam particle. For each allowed set of incoming flavours, the full matrix-element expression is put together, using the common pieces and the flavour-dependent couplings. This is multiplied by the common factors and the parton-distribution weights to obtain a cross-section weight. \item Each allowed flavour combination is stored as a separate entry in a table, together with its weight. In addition, a summed weight is calculated. \item The phase-space point is kept or rejected, according to a comparison of the summed weight with the maximum weight obtained at initialization. Also the cross-section Monte Carlo integration is based on the summed weight. \item If the point is retained, one of the allowed flavour combinations is picked according to the relative weights stored in the full table. \end{Enumerate} Generally, the flavours of the final state are either completely specified by those of the initial state, e.g. as in $\q\g \to \q\g$, or completely decoupled from them, e.g. as in $\f\fbar \to \Z^0 \to \f'\fbar'$. In neither case need therefore the final-state flavours be specified in the cross-section calculation. It is only necessary, in the latter case, to include an overall weight factor, which takes into account the summed contribution of all final states that are to be simulated. For instance, if only the process $\Z^0 \to \ee$ is studied, the relevant weight factor is simply $\Gamma_{\e\e} / \Gamma_{\mrm{tot}}$. Once the kinematics and the incoming flavours have been selected, the outgoing flavours can be picked according to the appropriate relative probabilities. In some processes, such as $\g\g \to \g\g$, several different colour flows are allowed, each with its own kinematical dependence of the matrix-element weight, see section \ref{sss:QCDjetclass}. Each colour flow is then given as a separate entry in the table mentioned above, i.e. in total an entry is characterized by the two incoming flavours, a colour-flow index, and the weight. For an accepted phase-space point, the colour flow is selected in the same way as the incoming flavours. The program can also allow the mixed generation of two or more completely different processes, such as $\f\fbar \to \Z^0$ and $\q\qbar \to \g\g$. In that case, each process is initialized separately, with its own set of coefficients $c_i$ and so on. The maxima obtained for the individual cross sections are all expressed in the same units, even when the dimensionality of the phase space is different. (This is because we always transform to a phase space of unit volume, $\int h_{\tau}(\tau) \, \d \tau \equiv 1$, etc.) The above generation scheme need therefore only be generalized as follows: \begin{Enumerate} \item One process is selected among the allowed ones, with a relative probability given by the maximum weight for this process. \item A phase-space point is found, using the distributions $h_{\tau}(\tau)$ and so on, optimized for this particular process. \item The total weight for the phase-space point is evaluated, again with `Jacobians', matrix elements and allowed incoming flavour combinations that are specific to the process. \item The point is retained with a probability given by the ratio of the actual to the maximum weight of the process. If the point is rejected, one has to go back to step 1 and pick a new process. \item Once a phase-space point has been accepted, flavours may be selected, and the event generated in full. \end{Enumerate} It is clear why this works: although phase-space points are selected among the allowed processes according to relative probabilities given by the maximum weights, the probability that a point is accepted is proportional to the ratio of actual to maximum weight. In total, the probability for a given process to be retained is therefore only proportional to the average of the actual weights, and any dependence on the maximum weight is gone. \subsection{$2 \to 3$ and $2 \to 4$ Processes} The {\Py} machinery to handle $2 \to 1$ and $2 \to 2$ processes is fairly sophisticated and generic. The same cannot be said about the generation of hard scattering processes with more than two final-state particles. The number of phase-space variables is larger, and it is therefore more difficult to find and transform away all possible peaks in the cross section by a suitably biased choice of phase-space points. In addition, matrix-element expressions for $2 \to 3$ processes are typically fairly lengthy. Therefore {\Py} only contains a very limited number of $2 \to 3$ and $2 \to 4$ processes, and almost each process is a special case of its own. It is therefore less interesting to discuss details, and we only give a very generic overview. If the Higgs mass is not light, interactions among longitudinal $\W$ and $\Z$ gauge bosons are of interest. In the program, $2 \to 1$ processes such as $\W_{\mrm{L}}^+ \W_{\mrm{L}}^- \to \H^0$ and $2 \to 2$ ones such as $\W_{\mrm{L}}^+ \W_{\mrm{L}}^- \to \Z_{\mrm{L}}^0 \Z_{\mrm{L}}^0$ are included. The former are for use when the $\H^0$ still is reasonably narrow, such that a resonance description is applicable, while the latter are intended for high energies, where different contributions have to be added up. Since the program does not contain $\W_{\mrm{L}}$ or $\Z_{\mrm{L}}$ distributions inside hadrons, the basic hard scattering has to be convoluted with the $\q \to \q' \W_{\mrm{L}}$ and $\q \to \q \Z_{\mrm{L}}$ branchings, to yield effective $2 \to 3$ and $2 \to 4$ processes. However, it is possible to integrate out the scattering angles of the quarks analytically, as well as one energy-sharing variable \cite{Cha85}. Only after an event has been accepted are these other kinematical variables selected. This involves further choices of random variables, according to a separate selection loop. In total, it is therefore only necessary to introduce one additional variable in the basic phase-space selection, which is chosen to be $\hat{s}'$, the squared invariant mass of the full $2 \to 3$ or $2 \to 4$ process, while $\hat{s}$ is used for the squared invariant mass of the inner $2 \to 1$ or $2 \to 2$ process. The $y$ variable is coupled to the full process, since parton-distribution weights have to be given for the original quarks at $x_{1,2} = \sqrt{\tau'} \exp{(\pm y)}$. The $\hat{t}$ variable is related to the inner process, and thus not needed for the $2 \to 3$ processes. The selection of the $\tau' = \hat{s}'/s$ variable is done after $\tau$, but before $y$ has been chosen. To improve the efficiency, the selection is made according to a weighted phase space of the form $\int h_{\tau'}(\tau') \, \d \tau'$, where \begin{equation} h_{\tau'}(\tau') = \frac{c_1}{{\cal I}_1} \frac{1}{\tau'} + \frac{c_2}{{\cal I}_2} \, \frac{(1- \tau / \tau')^3}{\tau'^2} + \frac{c_3}{{\cal I}_3} \, \frac{1}{\tau' (1 - \tau')} ~, \end{equation} in conventional notation. The $c_i$ coefficients are optimized at initialization. The $c_3$ term, peaked at $\tau' \approx 1$, is only used for $\ee$ collisions. The choice of $h_{\tau'}$ is roughly matched to the longitudinal gauge-boson flux factor, which is of the form \begin{equation} \left( 1 + \frac{\tau}{\tau'} \right) \, \ln \left( \frac{\tau}{\tau'} \right) - 2 \left( 1 - \frac{\tau}{\tau'} \right) ~. \end{equation} For a light $\H$ the effective $\W$ approximation above breaks down, and it is necessary to include the full structure of the $\q \q' \to \q \q' \H^0$ (i.e. $\Z\Z$ fusion) and $\q \q' \to \q'' \q''' \H^0$ (i.e. $\W\W$ fusion) matrix elements. The $\tau'$, $\tau$ and $y$ variables are here retained, and selected according to standard procedures. The Higgs mass is represented by the $\tau$ choice; normally the $\H^0$ is so narrow that the $\tau$ distribution effectively collapses to a $\delta$ function. In addition, the three-body final-state phase space is rewritten as \begin{equation} \left( \prod_{i=3}^5 \frac{1}{(2 \pi)^3} \frac{\d^3 p_i}{2 E_i} \right) \, (2 \pi)^4 \delta^{(4)} (p_3 + p_4 + p_5 - p_1 - p_2) = \frac{1}{(2 \pi)^5} \, \frac{\pi^2}{4 \sqrt{\lambda_{\perp 34}}} \, \d p_{\perp 3}^2 \, \frac{\d \varphi_3}{2 \pi} \, \d p_{\perp 4}^2 \, \frac{\d \varphi_4}{2 \pi} \, \d y_5 ~, \end{equation} where $\lambda_{\perp 34} = (m_{\perp 34}^2 - m_{\perp 3}^2 - m_{\perp 4}^2)^2 - 4 m_{\perp 3}^2 m_{\perp 4}^2$. The outgoing quarks are labelled 3 and 4, and the outgoing Higgs 5. The $\varphi$ angles are selected isotropically, while the two transverse momenta are picked, with some foreknowledge of the shape of the $\W / \Z$ propagators in the cross sections, according to $h_{\perp} (\pT^2) \, \d \pT^2$, where \begin{equation} h_{\perp}(\pT^2) = \frac{c_1}{{\cal I}_1} + \frac{c_2}{{\cal I}_2} \, \frac{1}{m_R^2 + \pT^2} + \frac{c_3}{{\cal I}_3} \, \frac{1}{(m_R^2 + \pT^2)^2} ~, \end{equation} with $m_R$ the $\W$ or $\Z$ mass, depending on process, and $c_1 = c_2 = 0.05$, $c_3 = 0.9$. Within the limits given by the other variable choices, the rapidity $y_5$ is chosen uniformly. A final choice remains to be made, which comes from a twofold ambiguity of exchanging the longitudinal momenta of partons 3 and 4 (with minor modifications if they are massive). Here the relative weight can be obtained exactly from the form of the matrix element itself. No good phase-space choice was found for the process $\g \g \to \Z^0 \b \bbar$. This process is therefore not so easy to generate with {\Py}. What is currently done is to use the basic formalism of $2 \to 2$ processes, where the $\b + \bbar$ system is considered as an effective `resonance'. Two masses are then selected, the $\Z^0$ one according to eq. (\ref{pg:reshdist}) and the $\b + \bbar$ one according to $\d m^2 / m^2$. Both `decays' are selected isotropically in the respective rest frame, to give the final four fermions in terms of which the matrix element is given. In addition, $\tau$, $y$ and $z$ are selected according to the standard rules for $2 \to 2$ processes. \subsection{Resonance Decays} \label{ss:resdecay} Resonances can be made to decay in two different routines. One is the standard decay treatment (in {\tt LUDECY}) that can be used for any unstable particle, where decay channels are chosen according to fixed probabilities, and decay angles usually are picked isotropically in the rest frame of the resonance, see section \ref{ss:partdecays}. The more sophisticated treatment (in {\tt PYRESD}) is the default one for resonances produced in {\Py}, and is described here. The following are included in the list of resonances: $\Z^0$, $\W^{\pm}$, $\H^0$, $\Z'^0$, $\W'^{\pm}$, $\H'^0$, $\A^0$, $\H^{\pm}$, $\eta_{\mrm{tech}}^0$, $\L_{\Q}$, and $\R^0$. The top is also considered as a resonance if it is assumed to decay before it has time to fragment. Likewise for the fourth generation fermions. If the fourth generation is used to represent excited quarks and leptons, these are also considered to be resonances. \subsubsection{The decay scheme} In the beginning of the decay treatment, either one or two resonances may be present, the former represented by processes such as $\q \qbar' \to \W^+$ and $\q \g \to \W^+ \q'$, the latter by $\q \qbar \to \W^+ \W^-$. If the latter is the case, the decay of the two resonances is considered in parallel (unlike {\tt LUDECY}, where one particle at a time is made to decay). First the decay channel of each resonance is selected according to the relative weights $H_R^{(f)}$, as described above, evaluated at the actual mass of the resonance, rather than at the nominal one. Threshold factors are therefore fully taken into account, with channels automatically switched off below the threshold. Normally the masses of the decay products are well-defined, but e.g. in decays like $\H^0 \to \W^+ \W^-$ it is also necessary to select the decay product masses. This is done according to two Breit--Wigners of the type in eq. (\ref{pg:resshapethree}), multiplied by the threshold factor, which depends on both masses. Next the decay angles of the resonance are selected isotropically in its rest frame. Normally the full range of decay angles is available, but in $2 \to 1$ processes the decay angles of the original resonance may be restrained by user cuts, e.g. on the $\pT$ of the decay products. Based on the angles, the four-momenta of the decay products are constructed and boosted to the correct frame. As a rule, matrix elements are given with quark and lepton masses assumed vanishing. Therefore the four-momentum vectors constructed at this stage are actually massless for all quarks and leptons. The matrix elements may now be evaluated. For a process such as $\q \qbar \to \W^+ \W^- \to \e^+ \nu_{\e} \mu^- \br{\nu}_{\mu}$, the matrix element is a function of the four-momenta of the two incoming fermions and of the four outgoing ones. An upper limit for the event weight can be constructed from the cross section for the basic process $\q \qbar \to \W^+ \W^-$, as already used to select the two $\W$ momenta. If the weighting fails, new resonance decay angles are picked and the procedure is iterated until acceptance. Based on the accepted set of angles, the correct decay product four-momenta are constructed, including previously neglected fermion masses. Quarks and, optionally, leptons are allowed to radiate, using the standard final-state showering machinery, with maximum virtuality given by the resonance mass. In some decays new resonances are produced, and these are then subsequently allowed to decay. Only one resonance pair is considered at a time, i.e. it is not possible to include correlations which involve the simultaneous decay of three or more resonances. This is in fact all that is currently needed: in a process like $\q\qbar \to \Z^0 \H^0 \to \Z^0 \W^+ \W^- \to 6$ fermions, the spinless nature of the $\H^0$ ensures that the $\W^{\pm}$ decays are decoupled from that of the $\Z^0$ (but not from each other). \subsubsection{Cross-section considerations} \label{sss:resdecaycross} The cross section for a process which involves the production of one or several resonances is always reduced to take into account channels not allowed by user flags. This is trivial for a single $s$-channel resonance, cf. eq. (\ref{pg:Hinoutsym}), but can also be included approximately if several layers of resonance decays are involved. At initialization, the ratio between the user-allowed width and the nominally possible one is evaluated and stored, starting from the lightest resonances and moving upwards. As an example, one first finds the reduction factors for $\W^+$ and for $\W^-$ decays, which need not be the same if e.g. $\W^+$ is allowed to decay only to quarks and $\W^-$ only to leptons. These factors enter together as a weight for the $\H^0 \to \W^+ \W^-$ channel, which is thus reduced in importance compared with other possible Higgs decay channels. This is also reflected in the weight factor of the $\H^0$ itself, where some channels are open in full, others completely closed, and finally some (like the one above) open but with reduced weight. Finally, the weight for the process $\q\qbar \to \Z^0 \H^0$ is evaluated as the product of the $\Z^0$ weight factor and the $\H^0$ one. The standard cross section of the process is multiplied with this weight. Since the restriction on allowed decay modes is already included in the hard process cross section, mixing of different event types is greatly simplified, and the selection of decay channel chains is straightforward. There is a price to be paid, however. The reduction factors evaluated at initialization all refer to resonances at their nominal masses. For instance, the $\W$ reduction factor is evaluated at the nominal $\W$ mass, even when that factor is used, later on, in the description of the decay of a 120 GeV Higgs, where at least one $\W$ would be produced below this mass. We know of no case where this approximation has any serious consequences, however. The weighting procedure works because the number of resonances to be produced, directly or in subsequent decays, can be derived recursively already from the start. It does not work for particles which could also be produced at later stages, such as the parton-shower evolution and the fragmentation. For instance, $\D^0$ mesons can be produced fairly late in the event generation chain, in unknown numbers, and so weights could not be introduced to compensate, e.g. for the forcing of decays only into $\pi^+ \K^-$. For similar reasons the top is only considered as a resonance if it is not allowed to hadronize; see discussion in section \ref{sss:heavflavclass}. One should note that this reduction factor is separate from the description of the resonance shape itself, where the full width of the resonance has to be used. This width is based on the sum of all possible decay modes, not just the simulated ones. {\Py} does allow the possibility to change also the underlying physics scenario, e.g. to include the decay of a $\Z^0$ into a fourth-generation neutrino. Normally the evaluation of the reduction factors is straightforward. However, for decays into a pair of equal or charge-conjugate resonances, such as $\Z^0 \Z^0$ or $\W^+ \W^-$, it is possible to pick combinations in such a way that the weight of the pair does not factorize into a product of the weight of each resonance itself. To be precise, any decay channel can be given seven different status codes: \begin{Itemize} \item $-1$: a non-existent decay mode, completely switched off and of no concern to us; \item 0: an existing decay channel, which is switched off; \item 1: a channel which is switched on; \item 2: a channel switched on for particles, but off for antiparticles; \item 3: a channel switched on for antiparticles, but off for particles; \item 4: a channel switched on for one of the resonances, but not for both; \item 5: a channel switched on for the other of the resonances, but not for both. \end{Itemize} The meaning of possibilities 4 and 5 is exemplified by the statement `in a $\W^+\W^-$ pair, one $\W$ decays hadronically and the other leptonically', which thus covers the cases where either $\W^+$ or $\W^-$ decays hadronically. Neglecting non-existing channels, each channel belongs to either of the classes above. If we denote the total branching ratio into channels of type $i$ by $r_i$, this then translates into the requirement $r_0 + r_1 + r_2 + r_3 + r_4 + r_5 = 1$ (unless some channels already come with a reduced weight because of restrictions on subsequent decay chains, in which case the sum is less than unity). For a single particle the weight factor is $r_1 + r_2 + r_4$, and for a single antiparticle $r_1 + r_3 + r_4$. For a pair of identical resonances, the joint weight is instead \begin{equation} (r_1 + r_2)^2 + 2 (r_1 + r_2) (r_4 + r_5) + 2 r_4 r_5 ~, \end{equation} and for a resonance--antiresonance pair \begin{equation} (r_1 + r_2)(r_1 + r_3) + (2 r_1 + r_2 + r_3) (r_4 + r_5) + 2 r_4 r_5 ~. \label{eq:WWallchancomb} \end{equation} \subsection{Nonperturbative Processes} \label{ss:nonpertproc} A few processes are not covered by the discussion so far. These are the ones that depend on the details of hadronic wave functions, and therefore are not strictly calculable perturbatively (although perturbation theory may often provide some guidance). What we have primarily in mind is elastic scattering, diffractive scattering and low-$\pT$ `minimum-bias' events in hadron--hadron collisions, but one can also find corresponding processes in $\gamma \p$ and $\gamma \gamma$ interactions. The description of these processes is rather differently structured from that of the other ones, as is explained below. Models for `minimum-bias' events are discussed in detail in section \ref{ss:multint}, to which we refer for details on this part of the program. \subsubsection{Hadron--hadron interactions} In hadron--hadron interactions, the total hadronic cross section for $AB \to$ anything, $\sigma^{AB}_{\mrm{tot}}$, is calculated using the parametrization of Donnachie and Landshoff \cite{Don92}. In this approach, each cross section appears as the sum of one pomeron term and one reggeon one \begin{equation} \sigma^{AB}_{\mrm{tot}}(s) = X^{AB} \, s^{\epsilon} + Y^{AB} \, s^{-\eta} ~, \label{pg:sigtotpomreg} \end{equation} where $s = E_{\mrm{cm}}^2$. The powers $\epsilon = 0.0808$ and $\eta = 0.4525$ are expected to be universal, whereas the coefficients $X^{AB}$ and $Y^{AB}$ are specific to each initial state. (In fact, the high-energy behaviour given by the pomeron term is expected to be the same for particle and antiparticle interactions, i.e. $X^{\br{A}B} = X^{AB}$.) Parametrizations not provided in \cite{Don92} have been calculated in the same spirit, making use of quark counting rules \cite{Sch93a}. The total cross section is subdivided according to \begin{equation} \sigma^{AB}_{\mrm{tot}}(s) = \sigma^{AB}_{\mrm{el}}(s) + \sigma^{AB}_{\mrm{sd}(XB)}(s) + \sigma^{AB}_{\mrm{sd}(AX)}(s) + \sigma^{AB}_{\mrm{dd}}(s) + \sigma^{AB}_{\mrm{nd}}(s) ~. \label{pg:sigtotsplit} \end{equation} Here `el' is the elastic process $AB \to AB$, `sd$(XB)$' the single diffractive $AB \to XB$, `sd$(AX)$' the single diffractive $AB \to AX$, `dd' the double diffractive $AB \to X_1 X_2$, and `nd' the non-diffractive ones. Higher diffractive topologies, such as central diffraction, are currently neglected. In the following, the elastic and diffractive cross sections and event characteristics are described, as given in the model by Schuler and Sj\"ostrand \cite{Sch93b,Sch93a}. The non-diffractive component is identified with the `minimum bias' physics already mentioned, a practical but not unambiguous choice. Its cross section is given by `whatever is left' according to eq. (\ref{pg:sigtotsplit}), and its properties are discussed in section \ref{ss:multint}. At not too large squared momentum transfers $t$, the elastic cross section can be approximated by a simple exponential fall-off. If one neglects the small real part of the cross section, the optical theorem then gives \begin{equation} \frac{\d\sigma_{\mrm{el}}}{\d t} = \frac{\sigma_{\mrm{tot}}^2}{16 \pi} \, \exp(B_{\mrm{el}} t) ~, \end{equation} and $\sigma_{\mrm{el}} = \sigma_{\mrm{tot}}^2 / 16 \pi B_{\mrm{el}}$. The elastic slope parameter is parametrized by \begin{equation} B_{\mrm{el}} = B^{AB}_{\mrm{el}}(s) = 2 b_A + 2 b_B + 4 s^{\epsilon} -4.2 ~, \end{equation} with $s$ given in units of GeV and $B_{\mrm{el}}$ in GeV$^{-2}$. The constants $b_{A,B}$ are $b_{\p} = 2.3$, $b_{\pi,\rho,\omega,\phi} = 1.4$, $b_{\J/\psi} = 0.23$. The increase of the slope parameter with c.m. energy is faster than the logarithmically one conventionally assumed; that way the ratio $\sigma_{\mrm{el}} / \sigma_{\mrm{tot}}$ remains well-behaved at large energies. The diffractive cross sections are given by \begin{eqnarray} \frac{\d\sigma_{\mrm{sd}(XB)}(s)}{\d t \, \d M^2} & = & \frac{g_{3\pomeron}}{16\pi} \, \beta_{A\pomeron} \, \beta_{B\pomeron}^2 \, \frac{1}{M^2} \, \exp(B_{\mrm{sd}(XB)}t) \, F_{\mrm{sd}} ~, \nonumber \\ \frac{\d\sigma_{\mrm{sd}(AX)}(s)}{\d t \, \d M^2} & = & \frac{g_{3\pomeron}}{16\pi} \, \beta_{A\pomeron}^2 \, \beta_{B\pomeron} \, \frac{1}{M^2} \, \exp(B_{\mrm{sd}(AX)}t) \, F_{\mrm{sd}} ~, \nonumber \\ \frac{\d\sigma_{\mrm{dd}}(s)}{\d t \, \d M_1^2 \, \d M_2^2} & = & \frac{g_{3\pomeron}^2}{16\pi} \, \beta_{A\pomeron} \, \beta_{B\pomeron} \, \frac{1}{M_1^2} \, \frac{1}{M_2^2} \, \exp(B_{\mrm{dd}}t) \, F_{\mrm{dd}} ~. \end{eqnarray} The couplings $\beta_{A\pomeron}$ are related to the pomeron term $X^{AB} s^{\epsilon}$ of the total cross section parametrization, eq. (\ref{pg:sigtotpomreg}). Picking a reference scale $\sqrt{s_{\mrm{ref}}} = 20$ GeV, the couplings are given by $\beta_{A\pomeron}\beta_{B\pomeron} = X^{AB} \, s_{\mrm{ref}}^{\epsilon}$. The triple-pomeron coupling is determined from single-diffractive data to be $g_{3\pomeron} \approx 0.318$ mb$^{1/2}$; within the context of the formulae in this section. The spectrum of diffractive masses $M$ is taken to begin 0.28 GeV $\approx 2 m_{\pi}$ above the mass of the respective incoming particle and extend to the kinematical limit. The simple $\d M^2 / M^2$ form is modified by the mass-dependence in the diffractive slopes and in the $F_{\mrm{sd}}$ and $F_{\mrm{dd}}$ factors. The slope parameters are assumed to be \begin{eqnarray} B_{\mrm{sd}(XB)}(s) & = & 2b_B + 2\alpha' \ln\left(\frac{s}{M^2}\right) ~, \nonumber \\ B_{\mrm{sd}(AX)}(s) & = & 2b_A + 2\alpha' \ln\left(\frac{s}{M^2}\right) ~, \nonumber \\ B_{\mrm{dd}}(s) & = & 2\alpha' \ln\left(e^4 + \frac{s s_0}{M_1^2 M_2^2} \right) ~. \end{eqnarray} Here $\alpha' = 0.25$ GeV$^{-2}$ and conventionally $s_0$ is picked as $s_0 = 1 / \alpha'$. The term $e^4$ in $B_{\mrm{dd}}$ is added by hand to avoid a breakdown of the standard expression for large values of $M_1^2 M_2^2$. The $b_{A,B}$ terms protect $B_{\mrm{sd}}$ from breaking down; however a minimum value of 2 is still explicitly required for $B_{\mrm{sd}}$, which comes into play e.g. for a $\Jpsi$ state (as part of a VMD photon beam). The kinematical range in $t$ depends on all the masses of the problem. In terms of the scaled variables $\mu_1 = m_A^2/s$, $\mu_2 = m_B^2/s$, $\mu_3 = M_{(1)}^2/s$ ($=m_A^2/s$ when $A$ scatters elastically), $\mu_4 = M_{(2)}^2/s$ ($=m_B^2/s$ when $B$ scatters elastically), and the combinations \begin{eqnarray} C_1 & = & 1 - (\mu_1 + \mu_2 + \mu_3 + \mu_4) + (\mu_1 - \mu_2) (\mu_3 - \mu_4) ~, \nonumber \\ C_2 & = & \sqrt{(1 - \mu_1 -\mu_2)^2 - 4 \mu_1 \mu_2} \, \sqrt{(1 - \mu_3 - \mu_4)^2 - 4 \mu_3 \mu_4} ~, \nonumber \\ C_3 & = & (\mu_3 - \mu_1) (\mu_4 - \mu_2) + (\mu_1 + \mu_4 - \mu_2 - \mu_3) (\mu_1 \mu_4 - \mu_2 \mu_3) ~, \end{eqnarray} one has $t_{\mmin} < t < t_{\mmax}$ with \begin{eqnarray} t_{\mmin} & = & - \frac{s}{2} (C_1 + C_2) ~, \nonumber \\ t_{\mmax} & = & - \frac{s}{2} (C_1 - C_2) = - \frac{s}{2} \, \frac{4C_3}{C_1 + C_2} = \frac{s^2 C_3}{t_{\mmin}} ~. \end{eqnarray} The Regge formulae above for single- and double-diffractive events are supposed to hold in certain asymptotic regions of the total phase space. Of course, there will be diffraction also outside these restrictive regions. Lacking a theory which predicts differential cross sections at arbitrary $t$ and $M^2$ values, the Regge formulae are used everywhere, but fudge factors are introduced in order to obtain `sensible' behaviour in the full phase space. These factors are: \begin{eqnarray} F_{\mrm{sd}} & = & \left( 1 - \frac{M^2}{s} \right) \left( 1 + \frac{c_{\mrm{res}} \, M_{\mrm{res}}^2} {M_{\mrm{res}}^2 + M^2} \right) ~, \nonumber \\ F_{\mrm{dd}} & = & \left( 1 - \frac{\left( M_1 + M_2 \right)^2}{s} \right) \left( \frac{s\, m_{\p}^2}{ s\, m_{\p}^2 + M_1^2\, M_2^2} \right) \nonumber \\ & \times & \left( 1 + \frac{c_{\mrm{res}} \, M_{\mrm{res}}^2} {M_{\mrm{res}}^2 + M_1^2} \right) \left( 1 + \frac{c_{\mrm{res}} \, M_{\mrm{res}}^2} {M_{\mrm{res}}^2 + M_2^2} \right) ~. \end{eqnarray} The first factor in either expression suppresses production close to the kinematical limit. The second factor in $F_{dd}$ suppresses configurations where the two diffractive systems overlap in rapidity space. The final factors give an enhancement of the low-mass region, where a resonance structure is observed in the data. Clearly a more detailed modelling would have to be based on a set of exclusive states rather than on this smeared-out averaging procedure. A reasonable fit to $\p\p / \pbar\p$ data is obtained for $c_{\mrm{res}} = 2$ and $M_{\mrm{res}} = 2$~GeV, for an arbitary particle $A$ which is diffractively excited we use $M_{\mrm{res}}^A = m_A - m_{\p} + 2$~GeV. The diffractive cross-section formulae above have been integrated for a set of c.m. energies, starting at 10 GeV, and the results have been parametrized. The form of these parametrizations is given in ref. \cite{Sch93b}, with explicit numbers for the $\p\p/\pbar\p$ case. {\Py} also contains similar parametrizations for $\pi\p$ (assumed to be same as $\rho\p$ and $\omega\p$), $\phi\p$, $\Jpsi\p$, and $\pi\pi$ ($\rho\rho$ etc.). The processes above do not obey the ordinary event mixing strategy. First of all, since their total cross sections are known, it is possible to pick the appropriate process from the start, and then remain with that choice. In other words, if the selection of kinematical variables fails, one would not go back and pick a new process, the way it was done in section \ref{sss:mixingproc}. Second, it is not possible to impose any cuts or restrain allowed incoming or outgoing flavours: if not additional information were to be provided, it would make the whole scenario ill-defined. Third, it is not recommended to mix generation of these processes with that of any of the other ones: normally the other processes have so small cross sections that they would almost never be generated anyway. (We here exclude the cases of `underlying events' and `pile-up events', where mixing is provided for, and even is a central part of the formalism, see sections \ref{ss:multint} and \ref{ss:pileup}.) Once the cross-section parametrizations has been used to pick one of the processes, the variables $t$ and $M$ are selected according to the formulae given above. A $\rho^0$ formed by $\gamma \to \rho^0$ in elastic or diffractive scattering is polarized, and therefore its decay angular distribution in $\rho^0 \to \pi^+ \pi^-$ is taken to be proportional to $\sin^2 \theta$, where the reference axis is given by the $\rho^0$ direction of motion. A light diffractive system, with a mass less than 1 GeV above the mass of the incoming particle, is allowed to decay isotropically into a two-body state. Single-resonance diffractive states, such as a $\Delta^+$, are therefore not explicily generated, but are assumed described in an average, smeared-out sense. A more massive diffractive system is subsequently treated as a string with the quantum numbers of the original hadron. Since the exact nature of the pomeron exchanged between the hadrons is unknown, two alternatives are included. In the first, the pomeron is assumed to couple to (valence) quarks, so that the string is stretched directly between the struck quark and the remnant diquark (antiquark) of the diffractive state. In the second, the interaction is rather with a gluon, giving rise to a `hairpin' configuration in which the string is stretched from a quark to a gluon and then back to a diquark (antiquark). Both of these scenarios could be present in the data; the default choice is to mix them in equal proportions. There is experimental support for more complicated scenarios \cite{Ing85}, wherein the pomeron has a partonic substructure, which e.g. can lead to high-$\pT$ jet production in the diffractive system. The full machinery, wherein a pomeron spectrum is folded with a pomeron-proton hard interaction, is not available in {\Py}. \subsubsection{Photoproduction} \label{sss:photoprod} The photoproduction part is still under active development. Currently only interactions between a hadron and a real photon have been studied in detail. In the future it is hoped to add $\gamma\gamma$ physics and interactions of mildly virtual photon (the transition region between real photons and deep inelastic scattering). The total $\gamma\p$ cross section can again be parametrized in a form like eq. (\ref{pg:sigtotpomreg}), which is not so obvious since the photon has more complicated structure than an ordinary hadron. In fact, the structure is still not so well understood. The model we outline is the one studied by Schuler and Sj\"ostrand \cite{Sch93,Sch93a}. In this model the physical photon is represented by \begin{equation} | \gamma \rangle = \sqrt{Z_3} \, | \gamma_B \rangle + \sum_{V=\rho^0,\omega,\phi,\Jpsi} \frac{e}{f_V} \, | V \rangle + \frac{e}{f_{\q\qbar}} \, | \q\qbar \rangle + \sum_{\ell=\e,\mu,\tau} \frac{e}{f_{\ell\ell}} \, | \ell^+ \ell^- \rangle ~. \label{pg:gammadecompo} \end{equation} By virtue of this superposition, one is led to a model of $\gamma\p$ interactions, where three different kinds of events may be distinguished: \begin{Itemize} \item Direct events, wherein the bare photon $| \gamma_B \rangle$ interacts directly with a parton from the proton. The process is perturbatively calculable, and no photon structure function is involved. The typical event structure is two high-$\pT$ jets and a proton remnant, while the photon does not leave behind any remnant. \item VMD events, in which the photon fluctuates into a vector meson, predominantly a $\rho^0$. All the event classes known from ordinary hadron--hadron interactions may thus occur here, such as elastic, diffractive, low-$\pT$ and high-$\pT$ events. For the latter, one may define a (VMD) photon structure function, and the photon also leaves behind a beam remnant. This remnant is smeared in transverse momentum by a typical `primordial $k_{\perp}$' of a few hundred MeV. \item Anomalous events, in which the photon fluctuates into a $\q\qbar$ pair of larger virtuality than in the VMD class. This process is perturbatively calculable, as is the subsequent QCD evolution. It gives rise to the so-called anomalous part of the photon structure function, whence the name for the class. It is assumed that only high-$\pT$ events may occur. Either the $\q$ or the $\qbar$ plays the r\^ole of a beam remnant, but this remnant has a larger $\pT$ than in the VMD case, related to the virtuality of the $\gamma \leftrightarrow \q\qbar$ fluctuation. \end{Itemize} The $| \ell^+ \ell^- \rangle$ states can only interact strongly with partons inside the hadron at higher orders, and can therefore be neglected. In order that the above classification is smooth and free of double counting, one has to introduce scales that separate the three components. The main one is $p_0$, which separates the low-mass vector meson region from the high-mass $| \q\qbar \rangle$ one, $p_0 \approx m_{\phi}/2 \approx 0.5$ GeV. Since it is the same $\gamma\q\qbar$ vertex that is responsible for the bare $\gamma \p$ interactions, $p_0$ is also the lower cut-off of the photon--parton cross sections. In addition, a $\pTmin$ cut-off is needed to separate low-$\pT$ and high-$\pT$ physics; see section \ref{ss:multint}. As it turns out, somewhat different $\pTmin$ values are needed for the VMD and anomalous parts; at least qualitatively this can be understood in terms of different sizes of the wave functions. The VMD and anomalous events are together called resolved ones. In terms of high-$\pT$ jet production, the VMD and anomalous contributions can be combined into a total resolved one, and the same for parton-distribution functions. However, the two classes differ in the structure of the underlying event and in the appearance of soft processes. In terms of cross sections, eq.\ (\ref{pg:gammadecompo}) corresponds to \begin{equation} \sigma_{\mrm{tot}}^{\gamma\p}(s) = \sigma_{\mrm{dir}}^{\gamma\p}(s) + \sigma_{\mrm{VMD}}^{\gamma\p}(s) + \sigma_{\mrm{anom}}^{\gamma\p}(s) ~. \label{pg:gammacrossdecompo} \end{equation} The direct cross section is, to lowest order, the perturbative cross section for the two processes $\gamma\q \to \q\g$ and $\gamma\g \to \q\qbar$, with a lower cut-off $\pT > p_0$. Properly speaking, this should be multiplied by the $Z_3$ coefficient, \begin{equation} Z_3 = 1 - \sum_{V=\rho^0,\omega,\phi,\Jpsi} \left( \frac{e}{f_V} \right)^2 - \left( \frac{e}{f_{\q\qbar}} \right)^2 - \sum_{\ell=\e,\mu,\tau} \left( \frac{e}{f_{\ell\ell}} \right)^2 ~, \end{equation} but normally $Z_3$ is so close to unity as to make no difference. The VMD factor $(e/f_V)^2 = 4\pi\alphaem/f_V^2$ gives the probability for the transition $\gamma \to V$. The coefficients $f_V^2/4\pi$ are determined from data to be (with a non-negligible amount of uncertainty) 2.20 for $\rho^0$, 23.6 for $\omega$, 18.4 for $\phi$ and 11.5 for $\Jpsi$. Together these numbers imply that the photon can be found in a VMD state about 0.4\% of the time, dominated by the $\rho^0$ contribution. All the properties of the VMD interactions can be obtained by appropriately scaling down $V\p$ physics predictions. Thus the whole machinery developed in the previous subsection for hadron--hadron interactions is directly applicable. Also parton distributions of the VMD component inside the photon are obtained by suitable rescaling. The contribution from the `anomalous' high-mass fluctuations depends on the typical scale $\mu$ of the interaction \begin{equation} \left( \frac{e}{f_{\q\qbar}} \right)^2 \approx \frac{\alphaem}{2\pi} \, \frac{N_C}{3} \, \left( 2 \sum_{\q} e_{\q}^2 \right) \ln\left( \frac{\mu^2}{p_0^2} \right) ~, \label{anomintfirst} \end{equation} where $N_C = 3$ and $\q$ runs over the quarks that can be taken massless compared with $\mu$. The logarithmic increase with $\mu$ implies that the anomalous contribution to the total photoproduction cross section ($\mu \sim m_V$) is less important than that to high-$\pT$ jet production ($\mu \sim \pT$). To first approximation, therefore only perturbative jet production above some $\pTmin$ scale is considered. This includes the standard QCD parton--parton scattering processes, with anomalous-photon parton distributions that are fully perturbatively calculable \cite{Sch93c}. In order to satisfy the equality in eq. (\ref{pg:gammacrossdecompo}), with the total cross section known and the direct and VDM contributions already fixed, a behaviour roughly like \begin{equation} \pTmin^{\mrm{anom}}(s) = 1.5~\mrm{GeV} + 0.0035 \sqrt{s} \end{equation} is needed over the HERA energy range. \clearpage \section{Physics Processes in PYTHIA} \label{s:pytproc} In this section we enumerate the physics processes that are available in {\Py}, introducing the ISUB code that can be used to select desired processes. A number of comments are made about the physics scenarios involved, in particular with respect to underlying assumptions and domain of validity. The section closes with a survey of interesting processes by machine. \subsection{The Process Classification Scheme} \label{ss:ISUBcode} A wide selection of fundamental $2 \to 1$ and $2 \to 2$ tree processes of the Standard Model (electroweak and strong) has been included in {\Py}, and slots are provided for many more, not yet implemented. In addition, a few `minimum-bias'-type processes (like elastic scattering), loop graphs, box graphs, $2 \to 3$ tree graphs and some non-Standard Model processes are included. The classification is not always unique. A process that proceeds only via an $s$-channel state is classified as a $2 \to 1$ process (e.g. $\q \qbar \to \gammaZ \to \ee$), but a $2 \to 2$ cross section may well have contributions from $s$-channel diagrams ($\g \g \to \g \g$ obtains contributions from $\g \g \to \g^* \to \g \g$). Also, in the program, $2 \to 1$ and $2 \to 2$ graphs may sometimes be folded with two $1 \to 2$ splittings to form effective $2 \to 3$ or $2 \to 4$ processes ($\W^+ \W^- \to \H^0$ is folded with $\q \to \q'' \W^+$ and $\q' \to \q''' \W^-$ to give $\q \q' \to \q'' \q''' \H^0$). It is possible to select a combination of subprocesses to simulate, and also afterwards to know which subprocess was actually selected in each event. For this purpose, all subprocesses are numbered according to an ISUB code. The list of possible codes is given in Tables \ref{t:procone}, \ref{t:proctwo}, \ref{t:procthree} and \ref{t:procfour}. Only processes marked with a `+' sign in the first column have been implemented in the program to date. Although ISUB codes were originally designed in a logical fashion, we must admit that subsequent developments of the program have tended to obscure the structure. For instance, the process numbers for Higgs production are spread out, in part as a consequence of the original classification, in part because further production mechanisms have been added one at a time, in whatever free slots could be found. At some future date the subprocess list will therefore be reorganized. In the thematic descriptions that follow the main tables, the processes of interest are repeated in a more logical order. If you want to look for a specific process, it will be easier to find it there. \begin{table}[pt] \caption{Subprocess codes, part 1. First column is `+' for processes implemented and blank for those that are only foreseen. Second is the subprocess number {\tt ISUB}, and third the description of the process. The final column gives references from which the cross sections have been obtained. See text for further information. \protect\label{t:procone} } \begin{center} \begin{tabular}{|c|r|l|l|@{\protect\rule{0mm}{\tablinsep}}} \hline In & No. & Subprocess & Reference \\ \hline & & a) $2 \to 1$, tree & \\ + & 1 & $\f_i \fbar_i \to \gammaZ$ & \cite{Eic84} \\ + & 2 & $\f_i \fbar_j \to \W^+$ & \cite{Eic84} \\ + & 3 & $\f_i \fbar_i \to \H^0$ & \cite{Eic84} \\ & 4 & $\gamma \W^+ \to \W^+$ & \\ + & 5 & $\Z^0 \Z^0 \to \H^0$ & \cite{Eic84,Cha85} \\ & 6 & $\Z^0 \W^+ \to \W^+$ & \\ & 7 & $\W^+ \W^- \to \Z^0$ & \\ + & 8 & $\W^+ \W^- \to \H^0$ & \cite{Eic84,Cha85} \\ \hline & & b) $2 \to 2$, tree & \\ + & 10 & $\f_i \f_j \to \f_i \f_j$ (QFD) & \cite{Ing87b} \\ + & 11 & $\f_i \f_j \to \f_i \f_j$ (QCD) & \cite{Com77,Ben84,Eic84} \\ + & 12 & $\f_i \fbar_i \to \f_k \fbar_k$ & \cite{Com77,Ben84,Eic84} \\ + & 13 & $\f_i \fbar_i \to \g \g$ & \cite{Com77,Ben84} \\ + & 14 & $\f_i \fbar_i \to \g \gamma$ & \cite{Hal78,Ben84} \\ + & 15 & $\f_i \fbar_i \to \g \Z^0$ & \cite{Eic84} \\ + & 16 & $\f_i \fbar_j \to \g \W^+$ & \cite{Eic84} \\ & 17 & $\f_i \fbar_i \to \g \H^0$ & \\ + & 18 & $\f_i \fbar_i \to \gamma \gamma$ & \cite{Ber84} \\ + & 19 & $\f_i \fbar_i \to \gamma \Z^0$ & \cite{Eic84} \\ + & 20 & $\f_i \fbar_j \to \gamma \W^+$ & \cite{Eic84,Sam91} \\ & 21 & $\f_i \fbar_i \to \gamma \H^0$ & \\ + & 22 & $\f_i \fbar_i \to \Z^0 \Z^0$ & \cite{Eic84,Gun86} \\ + & 23 & $\f_i \fbar_j \to \Z^0 \W^+$ & \cite{Eic84,Gun86} \\ + & 24 & $\f_i \fbar_i \to \Z^0 \H^0$ & \cite{Ber84} \\ + & 25 & $\f_i \fbar_i \to \W^+ \W^-$ & \cite{Eic84,Gun86} \\ + & 26 & $\f_i \fbar_j \to \W^+ \H^0$ & \cite{Eic84} \\ & 27 & $\f_i \fbar_i \to \H^0 \H^0$ & \\ + & 28 & $\f_i \g \to \f_i \g$ & \cite{Com77,Ben84} \\ + & 29 & $\f_i \g \to \f_i \gamma$ & \cite{Hal78,Ben84} \\ + & 30 & $\f_i \g \to \f_i \Z^0$ & \cite{Eic84} \\ + & 31 & $\f_i \g \to \f_k \W^+$ & \cite{Eic84} \\ & 32 & $\f_i \g \to \f_i \H^0$ & \\ + & 33 & $\f_i \gamma \to \f_i \g$ & \cite{Duk82} \\ + & 34 & $\f_i \gamma \to \f_i \gamma$ & \cite{Duk82} \\ + & 35 & $\f_i \gamma \to \f_i \Z^0$ & \cite{Gab86} \\ + & 36 & $\f_i \gamma \to \f_k \W^+$ & \cite{Gab86} \\ & 37 & $\f_i \gamma \to \f_i \H^0$ & \\ & 38 & $\f_i \Z^0 \to \f_i \g$ & \\ & 39 & $\f_i \Z^0 \to \f_i \gamma$ & \\ & 40 & $\f_i \Z^0 \to \f_i \Z^0$ & \\ \hline \end{tabular} \end{center} \end{table} \begin{table}[pt] \caption{Subprocess codes, part 2. First column is `+' for processes implemented and blank for those that are only foreseen. Second is the subprocess number {\tt ISUB}, and third the description of the process. The final column gives references from which the cross sections have been obtained. See text for further information. \protect\label{t:proctwo} } \begin{center} \begin{tabular}{|c|r|l|l|@{\protect\rule{0mm}{\tablinsep}}} \hline In & No. & Subprocess & Reference \\ \hline & & b) $2 \to 2$, tree (cont'd) & \\ & 41 & $\f_i \Z^0 \to \f_k \W^+$ & \\ & 42 & $\f_i \Z^0 \to \f_i \H^0$ & \\ & 43 & $\f_i \W^+ \to \f_k \g$ & \\ & 44 & $\f_i \W^+ \to \f_k \gamma$ & \\ & 45 & $\f_i \W^+ \to \f_k \Z^0$ & \\ & 46 & $\f_i \W^+ \to \f_k \W^+$ & \\ & 47 & $\f_i \W^+ \to \f_k \H^0$ & \\ & 48 & $\f_i \H^0 \to \f_i \g$ & \\ & 49 & $\f_i \H^0 \to \f_i \gamma$ & \\ & 50 & $\f_i \H^0 \to \f_i \Z^0$ & \\ & 51 & $\f_i \H^0 \to \f_k \W^+$ & \\ & 52 & $\f_i \H^0 \to \f_i \H^0$ & \\ + & 53 & $\g \g \to \f_k \fbar_k$ & \cite{Com77,Ben84} \\ + & 54 & $\g \gamma \to \f_k \fbar_k$ & \cite{Duk82} \\ & 55 & $\g \Z^0 \to \f_k \fbar_k$ & \\ & 56 & $\g \W^+ \to \f_k \fbar_l$ & \\ & 57 & $\g \H^0 \to \f_k \fbar_l$ & \\ + & 58 & $\gamma \gamma \to \f_k \fbar_k$ & \cite{Bar90} \\ & 59 & $\gamma \Z^0 \to \f_k \fbar_k$ & \\ & 60 & $\gamma \W^+ \to \f_k \fbar_l$ & \\ & 61 & $\gamma \H^0 \to \f_k \fbar_k$ & \\ & 62 & $\Z^0 \Z^0 \to \f_k \fbar_k$ & \\ & 63 & $\Z^0 \W^+ \to \f_k \fbar_l$ & \\ & 64 & $\Z^0 \H^0 \to \f_k \fbar_k$ & \\ & 65 & $\W^+ \W^- \to \f_k \fbar_k$ & \\ & 66 & $\W^+ \H^0 \to \f_k \fbar_l$ & \\ & 67 & $\H^0 \H^0 \to \f_k \fbar_k$ & \\ + & 68 & $\g \g \to \g \g$ & \cite{Com77,Ben84} \\ + & 69 & $\gamma \gamma \to \W^+ \W^-$ & \cite{Kat83} \\ + & 70 & $\gamma \W^+ \to \Z^0 \W^+$ & \cite{Kun87} \\ + & 71 & $\Z^0 \Z^0 \to \Z^0 \Z^0$ (longitudinal) & \cite{Abb87} \\ + & 72 & $\Z^0 \Z^0 \to \W^+ \W^-$ (longitudinal) & \cite{Abb87} \\ + & 73 & $\Z^0 \W^+ \to \Z^0 \W^+$ (longitudinal) & \cite{Dob91} \\ & 74 & $\Z^0 \H^0 \to \Z^0 \H^0$ & \\ & 75 & $\W^+ \W^- \to \gamma \gamma$ & \\ + & 76 & $\W^+ \W^- \to \Z^0 \Z^0$ (longitudinal) & \cite{Ben87b} \\ + & 77 & $\W^+ \W^{\pm} \to \W^+ \W^{\pm}$ (longitudinal) & \cite{Dun86,Bar90a} \\ & 78 & $\W^+ \H^0 \to \W^+ \H^0$ & \\ & 79 & $\H^0 \H^0 \to \H^0 \H^0$ & \\ + & 80 & $\q_i \gamma \to \q_k \pi^{\pm}$ & \cite{Bag82} \\ \hline \end{tabular} \end{center} \end{table} \begin{table}[pt] \caption{Subprocess codes, part 3. First column is `+' for processes implemented and blank for those that are only foreseen. Second is the subprocess number {\tt ISUB}, and third the description of the process. The final column gives references from which the cross sections have been obtained. See text for further information. \protect\label{t:procthree} } \begin{center} \begin{tabular}{|c|r|l|l|@{\protect\rule{0mm}{\tablinsep}}} \hline In & No. & Subprocess & Reference \\ \hline & & c) $2 \to 2$, tree, massive final quarks & \\ + & 81 & $\f_i \fbar_i \to \Q_k \Qbar_k$ & \cite{Com79} \\ + & 82 & $\g \g \to \Q_k \Qbar_k$ & \cite{Com79} \\ + & 83 & $\q_i \f_j \to \Q_k \f_l$ & \cite{Zer90} \\ + & 84 & $\g \gamma \to \Q_k \Qbar_k$ & \cite{Fon81} \\ + & 85 & $\gamma \gamma \to \F_k \Fbar_k$ & \cite{Bar90} \\ + & 86 & $\g \g \to \Jpsi \g$ & \cite{Bai83} \\ + & 87 & $\g \g \to \chi_{0 \c} \g$ & \cite{Gas87} \\ + & 88 & $\g \g \to \chi_{1 \c} \g$ & \cite{Gas87} \\ + & 89 & $\g \g \to \chi_{2 \c} \g$ & \cite{Gas87} \\ \hline & & d) `minimum bias' & \\ + & 91 & elastic scattering & \cite{Sch93b} \\ + & 92 & single diffraction ($AB \to XB$) & \cite{Sch93b} \\ + & 93 & single diffraction ($AB \to AX$) & \cite{Sch93b} \\ + & 94 & double diffraction & \cite{Sch93b} \\ + & 95 & low-$\pT$ production & \cite{Sjo87} \\ \hline & & e) $2 \to 1$, loop & \\ & 101 & $\g \g \to \Z^0$ & \\ + & 102 & $\g \g \to \H^0$ & \cite{Eic84} \\ + & 103 & $\gamma \gamma \to \H^0$ & \cite{Dre89} \\ \hline & & f) $2 \to 2$, box & \\ + & 110 & $\f_i \fbar_i \to \gamma \H^0$ & \cite{Ber85a} \\ + & 111 & $\f_i \fbar_i \to \g \H^0$ & \cite{Ell88} \\ + & 112 & $\f_i \g \to \f_i \H^0$ & \cite{Ell88} \\ + & 113 & $\g \g \to \g \H^0$ & \cite{Ell88} \\ + & 114 & $\g \g \to \gamma \gamma$ & \cite{Con71,Ber84,Dic88} \\ + & 115 & $\g \g \to \g \gamma$ & \cite{Con71,Ber84,Dic88} \\ & 116 & $\g \g \to \gamma \Z^0$ & \\ & 117 & $\g \g \to \Z^0 \Z^0$ & \\ & 118 & $\g \g \to \W^+ \W^-$ & \\ & 119 & $\gamma \gamma \to \g \g$ & \\ \hline & & g) $2 \to 3$, tree & \\ + & 121 & $\g \g \to \Q_k \Qbar_k \H^0$ & \cite{Kun84} \\ + & 122 & $\q_i \qbar_i \to \Q_k \Qbar_k \H^0$ & \cite{Kun84} \\ + & 123 & $\f_i \f_j \to \f_i \f_j \H^0$ ($\Z \Z$ fusion) & \cite{Cah84} \\ + & 124 & $\f_i \f_j \to \f_k \f_l \H^0$ ($\W^+\W^-$ fusion) & \cite{Cah84} \\ + & 131 & $\g \g \to \Z^0 \Q_k \br{\Q}_k$ & \cite{Eij90} \\ \hline \end{tabular} \end{center} \end{table} \begin{table}[pt] \caption{Subprocess codes, part 4. First column is `+' for processes implemented and blank for those that are only foreseen. Second is the subprocess number {\tt ISUB}, and third the description of the process. The final column gives references from which the cross sections have been obtained. See text for further information. \protect\label{t:procfour} } \begin{center} \begin{tabular}{|c|r|l|l|@{\protect\rule{0mm}{\tablinsep}}} \hline In & No. & Subprocess & Reference \\ \hline & & h) non-Standard Model, $2 \to 1$ & \\ + & 141 & $\f_i \fbar_i \to \gamma/\Z^0/\Z'^0$ & \cite{Alt89} \\ + & 142 & $\f_i \fbar_j \to \W'^+$ & \cite{Alt89} \\ + & 143 & $\f_i \fbar_j \to \H^+$ & \cite{Gun87} \\ + & 144 & $\f_i \fbar_j \to \R$ & \cite{Ben85a} \\ + & 145 & $\q_i \ell_j \to \L_{\Q}$ & \cite{Wud86} \\ + & 147 & $\d \g \to \d^*$ & \cite{Bau90} \\ + & 148 & $\u \g \to \u^*$ & \cite{Bau90} \\ + & 149 & $\g \g \to \eta_{\mrm{techni}}$ & \cite{Eic84,App92} \\ + & 151 & $\f_i \fbar_i \to \H'^0$ & \cite{Eic84} \\ + & 152 & $\g \g \to \H'^0$ & \cite{Eic84} \\ + & 153 & $\gamma \gamma \to \H'^0$ & \cite{Dre89} \\ + & 156 & $\f_i \fbar_i \to \A^0$ & \cite{Eic84} \\ + & 157 & $\g \g \to \A^0$ & \cite{Eic84} \\ + & 158 & $\gamma \gamma \to \A^0$ & \cite{Dre89} \\ \hline & & i) non-Standard Model, $2 \to 2$ and $2 \to 3$ & \\ + & 161 & $\f_i \g \to \f_k \H^+$ & \cite{Bar88} \\ + & 162 & $\q \g \to \ell \L_{\Q}$ & \cite{Hew88} \\ + & 163 & $\g \g \to \L_{\Q} \br{\L}_{\Q}$ & \cite{Hew88,Eic84} \\ + & 164 & $\q_i \qbar_i \to \L_{\Q} \br{\L}_{\Q}$ & \cite{Hew88} \\ + & 165 & $\f_i \fbar_i \to \f_k \fbar_k$ (via $\gammaZ$) & \cite{Eic84,Lan91} \\ + & 166 & $\f_i \fbar_j \to \f_k \fbar_l$ (via $\W^{\pm}$) & \cite{Eic84,Lan91} \\ + & 171 & $\f_i \fbar_i \to \Z^0 \H'^0$ & \cite{Eic84} \\ + & 172 & $\f_i \fbar_j \to \W^+ \H'^0$ & \cite{Eic84} \\ + & 173 & $\f_i \f_j \to \f_i \f_j \H'^0$ ($\Z \Z$ fusion) & \cite{Cah84} \\ + & 174 & $\f_i \f_j \to \f_k \f_l \H'^0$ ($\W^+\W^-$ fusion) & \cite{Cah84} \\ + & 176 & $\f_i \fbar_i \to \Z^0 \A^0$ & \cite{Eic84} \\ + & 177 & $\f_i \fbar_j \to \W^+ \A^0$ & \cite{Eic84} \\ + & 178 & $\f_i \f_j \to \f_i \f_j \A^0$ ($\Z \Z$ fusion) & \cite{Cah84} \\ + & 179 & $\f_i \f_j \to \f_k \f_l \A^0$ ($\W^+\W^-$ fusion) & \cite{Cah84} \\ + & 181 & $\g \g \to \Q_k \Qbar_k \H'^0$ & \cite{Kun84} \\ + & 182 & $\q_i \qbar_i \to \Q_k \Qbar_k \H'^0$ & \cite{Kun84} \\ + & 186 & $\g \g \to \Q_k \Qbar_k \A^0$ & \cite{Kun84} \\ + & 187 & $\q_i \qbar_i \to \Q_k \Qbar_k \A^0$ & \cite{Kun84} \\ \hline \end{tabular} \end{center} \end{table} In the following, $\f_i$ represents a fundamental fermion of flavour $i$, i.e. $\d$, $\u$, $\s$, $\c$, $\b$, $\t$, $\lrm$, $\hrm$, $\e^-$, $\nu_{\e}$, $\mu^-$, $\nu_{\mu}$, $\tau^-$, $\nu_{\tau}$, $\chi^-$ or $\nu_{\chi}$. A corresponding antifermion is denoted by $\fbar_i$. In several cases, some classes of fermions are explicitly excluded, since they do not couple to the $\g$ or $\gamma$ (no $\ee \to \g \g$, e.g.). When processes have only been included for quarks, while leptons might also have been possible, the notation $\q_i$ is used. A lepton is denoted by $\ell$; in a few cases neutrinos are also lumped under this heading. In processes where fermion masses are explicitly included in the matrix elements, an $\F$ is used to denote an arbitrary fermion and a $\Q$ a quark. Flavours appearing already in the initial state are denoted by indices $i$ and $j$, whereas new flavours in the final state are denoted by $k$ and $l$. Charge-conjugate channels are always assumed included as well (where separate), and processes involving a $\W^+$ also imply those involving a $\W^-$. Wherever $\Z^0$ is written, it is understood that $\gamma^*$ and $\gammaZ$ interference should be included as well (with possibilities to switch off either, if so desired). In some cases this is not fully implemented, see further below. Correspondingly, $\Z'^0$ denotes the complete set $\gamma^*/\Z^0/\Z'^0$ (or some subset of it). Thus the notation $\gamma$ is only used for a photon on the mass shell. In the last column of the tables below, references are given to works from which formulae have been taken. Sometimes these references are to the original works on the subject, sometimes only to the place where the formulae are given in the most convenient or accessible form, or where chance lead us. Apologies to all matrix-element calculators who are not mentioned. However, remember that this is not a review article on physics processes, but only a way for readers to know what is actually found in the program, for better or worse. In several instances, errata have been obtained from the authors. Often the formulae given in the literature have been generalized to include trivial radiative corrections, Breit--Wigner line shapes with $\hat{s}$-dependent widths (see section \ref{ss:kinemreson}), etc. The following sections contain some useful comments on the processes included in the program, grouped by physics interest rather than sequentially by ISUB or {\tt MSEL} code (see \ref{ss:PYswitchkin} for further information on the {\tt MSEL} code). The different ISUB and {\tt MSEL} codes that can be used to simulate the different groups are given. ISUB codes within brackets indicate the kind of processes that indirectly involve the given physics topic, although only as part of a larger whole. Some obvious examples, such as the possibility to produce jets in just about any process, are not spelled out in detail. The text at times contains information on which special switches or parameters are of particular interest to a given process. All these switches are described in detail in section \ref{ss:PYswitchpar}, but are alluded to here so as to provide a more complete picture of the possibilities available for the different subprocesses. However, the list of possibilities is certainly not exhausted by the text below. \subsection{QCD Processes} In this section we discuss scatterings exclusively between coloured partons --- a process like $\ee \to \gammaZ \to \q\qbar$ is also traditionally called a QCD event, but is here book-kept as $\gammaZ$ production. \subsubsection{QCD jets} \label{sss:QCDjetclass} {\tt MSEL} = 1, 2 \\ ISUB = \\ \begin{tabular}{rl@{\protect\rule{0mm}{\tablinsep}}} 11 & $\q_i \q_j \to \q_i \q_j$ \\ 12 & $\q_i \qbar_i \to \q_k \qbar_k$ \\ 13 & $\q_i \qbar_i \to \g \g$ \\ 28 & $\q_i \g \to \q_i \g$ \\ 53 & $\g \g \to \q_k \qbar_k$ \\ 68 & $\g \g \to \g \g$ \\ \end{tabular} No higher-order processes are explicitly included, nor any higher-order loop corrections to the $2 \to 2$ processes. However, by initial- and final-state QCD radiation, multijet events are being generated, starting from the above processes. The shower rate of multijet production is clearly uncertain by some amount, especially for well-separated jets. A string-based fragmentation scheme such as the Lund model needs cross sections for the different colour flows; these have been calculated in \cite{Ben84} and differ from the usual calculations by interference terms of the order $1/N_C^2$. By default, the standard QCD expressions for the differential cross sections are used. In this case, the interference terms are distributed on the various colour flows according to the pole structure of the terms. However, the interference terms can be excluded, by changing {\tt MSTP(34)} As an example, consider subprocess 28, $\q \g \to \q \g$. The total cross section for this process, obtained by summing and squaring the Feynman $\hat{s}$-, $\hat{t}$-, and $\hat{u}$-channel graphs, is \cite{Com77} \begin{equation} 2 \left( 1 - \frac{\hat{u}\hat{s}}{\hat{t}^2} \right) - \frac{4}{9} \left( \frac{\hat{s}}{\hat{u}} + \frac{\hat{u}}{\hat{s}} \right) - 1 ~. \end{equation} (An overall factor $\pi \alphas^2/\hat{s}^2$ is ignored.) Using the identity of the Mandelstam variables for the massless case, $\hat{s} + \hat{t} + \hat{u} = 0$, this can be rewritten as \begin{equation} \frac{\hat{s}^2 + \hat{u}^2}{\hat{t}^2} - \frac{4}{9} \left( \frac{\hat{s}}{\hat{u}} + \frac{\hat{u}}{\hat{s}} \right) ~. \end{equation} On the other hand, the cross sections for the two possible colour flows of this subprocess are \cite{Ben84} \begin{eqnarray} A: & & \frac{4}{9} \left( 2 \frac{\hat{u}^2}{\hat{t}^2} - \frac{\hat{u}}{\hat{s}} \right) ~; \nonumber \\ B: & & \frac{4}{9} \left( 2 \frac{\hat{s}^2}{\hat{t}^2} - \frac{\hat{s}}{\hat{u}} \right) ~. \end{eqnarray} Colour configuration $A$ is one in which the original colour of the $\q$ annihilates with the anticolour of the $\g$, the $\g$ colour flows through, and a new colour--anticolour is created between the final $\q$ and $\g$. In colour configuration $B$, the gluon anticolour flows through, but the $\q$ and $\g$ colours are interchanged. Note that these two colour configurations have different kinematics dependence. For {\tt MSTP(34)=0}, these are the cross sections actually used. The sum of the $A$ and $B$ contributions is \begin{equation} \frac{8}{9} \frac{\hat{s}^2 + \hat{u}^2}{\hat{t}^2} - \frac{4}{9} \left( \frac{\hat{s}}{\hat{u}} + \frac{\hat{u}}{\hat{s}} \right) ~. \end{equation} The difference between this expression and that of \cite{Com77}, corresponding to the interference between the two colour-flow configurations, is then \begin{equation} \frac{1}{9} \frac{\hat{s}^2 + \hat{u}^2}{\hat{t}^2} ~, \end{equation} which can be naturally divided between colour flows $A$ and $B$: \begin{eqnarray} A: & & \frac{1}{9} \frac{\hat{u}^2}{\hat{t}^2} ~; \nonumber \\ B: & & \frac{1}{9} \frac{\hat{s}^2}{\hat{t}^2} ~. \end{eqnarray} For {\tt MSTP(34)=1}, the standard QCD matrix element is therefore used, with the same relative importance of the two colour configurations as above. Similar procedures are followed also for the other QCD subprocesses. All the matrix elements in this group are for massless quarks (although final-state quarks are of course put on the mass shell). As a consequence, cross sections are divergent for $\pT \to 0$, and some kind of regularization is required. Normally you are expected to set the desired $\pTmin$ value in {\tt CKIN(3)}. The new flavour produced in the annihilation processes (ISUB = 12 and 53) is determined by the flavours allowed for gluon splitting into quark--antiquark; see switch {\tt MDME}. \subsubsection{Heavy flavours} \label{sss:heavflavclass} {\tt MSEL} = 4, 5, 6, 7, 8 \\ ISUB = \\ \begin{tabular}{rl@{\protect\rule{0mm}{\tablinsep}}} 81 & $\q_i \qbar_i \to \Q_k \Qbar_k$ \\ 82 & $\g \g \to \Q_k \Qbar_k$ \\ (83) & $\q_i \f_j \to Q_k f_l$ \\ \end{tabular} The matrix elements in this group differ from the corresponding ones in the group above in that they correctly take into account the quark masses. As a consequence, the cross sections are finite for $\pT \to 0$. It is therefore not necessary to introduce any special cuts. The two first processes that appear here are the dominant lowest-order graphs in hadron colliders --- a few other graphs will be mentioned later, such as process 83, which is important for a heavy top. The choice of flavour to produce is according to a hierarchy of options: \begin{Enumerate} \item if {\tt MSEL=4-8} then the flavour is set by the {\tt MSEL} value; \item else if {\tt MSTP(7)=1-8} then the flavour is set by the {\tt MSTP(7)} value; \item else the flavour is determined by the heaviest flavour allowed for gluon splitting into quark--antiquark; see switch {\tt MDME}. \end{Enumerate} Note that only one heavy flavour is allowed at a time; if more than one is turned on in {\tt MDME}, only the heaviest will be produced (as opposed to the case for ISUB = 12 and 53 above, where more than one flavour is allowed simultaneously). The lowest-order processes listed above just represent one source of heavy-flavour production. Heavy quarks can also be present in the parton distributions at the $Q^2$ scale of the hard interaction, leading to processes like $\Q \g \to \Q \g$, so-called flavour excitation, or they can be created by gluon splittings $\g \to \Q \Qbar$ in initial- or final-state shower evolution. In fact, as the c.m. energy is increased, these other processes gain in importance relative to the lowest-order production graphs above. As as example, only 10\% of the $\b$ production at LHC energies come from the lowest-order graphs. The figure is even smaller for charm, while it is at or above 50\% for top. At LHC energies, the specialized treatment described in this subsection is therefore only of interest for top (and potential fourth-generation quarks) --- the higher-order corrections can here be approximated by an effective $K$ factor, except possibly in some rare corners of phase space. For charm and bottom, on the other hand, it is necessary to simulate the full event sample (within the desired kinematics cuts), and then only keep those events with $\b/\c$ either from lowest-order production, or flavour excitation, or gluon splitting. Obviously this may be a time-consuming enterprise --- although the probability for a high-$\pT$ event to contain (at least) one charm or bottom pair is fairly large, most of these heavy flavours are carrying a small fraction of the total $\pT$ flow of the jets, and therefore do not survive normal experimental cuts. As an aside, it is not only for the lowest-order graphs that events may be generated with a guaranteed heavy-flavour content. One may also generate the flavour excitation process by itself, in the massless approximation, using ISUB = 28 and setting the {\tt KFIN} array appropriately. No trick exists to force the gluon splittings without introducing undesirable biases, however. The cross section for a heavy quark pair close to threshold can be modified according to the formulae of \cite{Fad90}, see {\tt MSTP(35)}. Here threshold effects due to $\Q\Qbar$ bound-state formation are taken into account in a smeared-out, average sense. Then the na\"{\i}ve cross section is multiplied by the squared wave function at the origin. In a colour-singlet channel this gives a net enhancement of the form \begin{equation} |\Psi^{(s)}(0)|^2 = \frac{X_{(s)}}{1 - \exp(- X_{(s)})} ~ , ~~~ \mrm{where}~ X_{(s)} = \frac{4}{3} \frac{\pi \alphas}{\beta} ~, \label{pp:threshenh} \end{equation} while in a colour octet channel there is a net suppression given by \begin{equation} |\Psi^{(8)}(0)|^2 = \frac{X_{(8)}}{\exp(- X_{(8)}) -1} ~ , ~~~ \mrm{where}~ X_{(8)} = \frac{1}{6} \frac{\pi \alphas}{\beta} ~. \label{pp:threshsup} \end{equation} The $\alphas$ factor in this expression is related to the energy scale of bound-state formation; it is selected independently from the one of the standard production cross section. The presence of a threshold factor affects the total rate and also kinematical distributions. Heavy flavours, i.e. top and fourth generation, are assumed to be so short-lived that they decay before they have time to hadronize. This means that the light quark in the decay $\Q \to \W^{\pm} \q$ inherits the colour of the heavy one. The new {\Py} description represents a change of philosophy compared to previous versions, formulated at a time when the top was thought to be much lighter than is believed currently. However, optionally the old description may still be used, where top hadrons are formed and these subsequently allowed to decay; see {\tt MSTP(48)} and {\tt MSTP(49)}. For event shapes the difference between the two time orderings normally has only marginal effects \cite{Sjo92a}. It should be noted that cross section calculations are different in the two cases. If the top (or the fourth generation fermion) is assumed short-lived, then it is treated like a resonance in the sense of section \ref{sss:resdecaycross}, i.e. the cross-section is reduced so as only to correspond to the channels left open by the user. This also includes the restrictions on secondary decays, i.e. on the decays of a $\W^+$ or a $\H^+$ produced in the top decay. If the top is allowed to form hadrons, no such reduction takes place. Branching ratios then have to be folded in by hand to get the correct cross sections. The logic behind this difference is that if hadronization takes place, one would be allowed e.g. to decay the $\T^0$ and $\T^+$ meson according to different branching ratios. But which $\T$ mesons are to be formed is not known at the top quark creation, so one could not weight for that. For a $\t$ quark which decays rapidly this ambiguity does not exist, and so a reduction factor can be introduced directly coupled to the $\t$ quark production process. This rule about cross-section calculations applies to all the processes explicitly set up to handle heavy flavour creation. In addition to the ones above, this means all the ones in Tables \ref{t:procone}--\ref{t:procfour} where the fermion final state is given as capital letters (`$\Q$' and `$\F$') and also flavours produced in resonance decays ($\Z^0$, $\W^{\pm}$, $\H^0$, etc., including processes 165 and 166). However, heavy flavours can also be produced in a process such as 31, $\q_i \g \to \q_k \W^{\pm}$, where $\q_k$ could be a top quark. In this case, the thrust of the description is clearly on light flavours --- the kinematics of the process is formulated in the massless fermion limit --- so any top production is purely incidental. Since here the choice of scattered flavour is only done at a later stage, the top branching ratios are not correctly folded in to the hard scattering cross section. So, for applications like these, it is not recommended to restrict the allowed top decay modes. Often one might like to get rid of the possibility of producing top together with light flavours. This can be done by switching off (i.e. setting {\tt MDME(I,1)=0}) the `channels' $\d \to \W^- \t$, $\s \to \W^- \t$, $\b \to \W^- \t$, $\g \to \t\tbar$ and $\gamma \to \t\tbar$. Also any heavy flavours produced by parton shower evolution would not be correctly weighted into the cross section. However, currently top production is switched off in both initial (see {\tt KFIN} array) and final (see {\tt MSTJ(45)}) state radiation. \subsubsection{J/$\psi$} \label{sss:Jpsiclass} ISUB = \\ \begin{tabular}{rl@{\protect\rule{0mm}{\tablinsep}}} 86 & $\g \g \to \Jpsi \g$ \\ 87 & $\g \g \to \chi_{0 \c} \g$ \\ 88 & $\g \g \to \chi_{1 \c} \g$ \\ 89 & $\g \g \to \chi_{2 \c} \g$ \\ \end{tabular} One may distinguish three main sources of $\Jpsi$ production. \begin{Enumerate} \item Decays of $\B$ mesons and baryons. \item Parton-shower evolution, wherein a $\c$ and a $\cbar$ quark produced in adjacent branchings (e.g. $\g \to \g \g \to \c \cbar \c \cbar$) turn out to have so small an invariant mass that the pair collapses to a single particle. \item Direct production, where a $\c$ quark loop gives a coupling between a set of gluons and a $\c\cbar$ bound state. Higher-lying states, like the $\chi_c$ ones, may subsequently decay to $\Jpsi$. \end{Enumerate} In this section are given the main processes for the third source, intended for applications at hadron colliders at non-vanishing transverse momenta --- in the limit of $\pT \to 0$ it is necessary to include a number of $2 \to 1$ processes and to regularize divergences in the $2 \to 2$ graphs above. The cross sections depend on wave function values at the origin, see {\tt PARP(38)} and {\tt PARP(39)}. A review of the physics issues involved may be found in \cite{Glo88} (note, however, that the choice of $Q^2$ scale is different in {\Py}). \subsubsection{Minimum bias} \label{sss:minbiasclass} {\tt MSEL} = 1, 2 \\ ISUB = \\ \begin{tabular}{rl@{\protect\rule{0mm}{\tablinsep}}} 91 & elastic scattering \\ 92 & single diffraction ($AB \to XB$) \\ 93 & single diffraction ($AB \to AX$) \\ 94 & double diffraction \\ 95 & low-$\pT$ production \\ \end{tabular} These processes are briefly discussed in section \ref{ss:nonpertproc}. Currently they are mainly intended for interactions between hadrons, although one may also consider $\gamma \p$ interactions in the option where the incoming photon is assumed resolved, {\tt MSTP(14)=1} or {\tt =2}. A possible extension to $\gamma \gamma$ interactions is not yet available. Uncertainties come from a number of sources, e.g. from the parametrizations of the various cross sections and slope parameters. In diffractive scattering, the structure of the selected hadronic system may be regulated with {\tt MSTP(101)}. No high-$\pT$ jet production in diffractive events is included so far. The subprocess 95, low-$\pT$ events, is somewhat unique in that no meaningful physical border-line to high-$\pT$ events can be defined. Even if the QCD $2 \to 2$ high-$\pT$ processes are formally switched off, some of the generated events will be classified as belonging to this group, with a $\pT$ spectrum of interactions to match the `minimum-bias' event sample. Only with the option {\tt MSTP(82)=0} will subprocess 95 yield strictly low-$\pT$ events, events which will then probably not be compatible with any experimental data. A number of options exist for the detailed structure of low-$\pT$ events, see in particular {\tt MSTP(81)} and {\tt MSTP(82)}. Further details on the model(s) for minimum-bias events are found in section \ref{ss:multint}. \subsection{Electroweak Gauge Bosons} This section covers the production and/or exchange of $\gamma$, $\Z^0$ and $\W^{\pm}$ gauge bosons, singly and in pairs. The topic of longitudinal gauge-boson scattering at high energies is deferred to the Higgs section, since the presence or absence of a Higgs here makes a big difference. \subsubsection{Prompt photon production} \label{sss:promptgammaclass} {\tt MSEL} = 10 \\ ISUB = \\ \begin{tabular}{rl@{\protect\rule{0mm}{\tablinsep}}} 14 & $\q_i \qbar_i \to \g \gamma$ \\ 18 & $\f_i \fbar_i \to \gamma \gamma$ \\ 29 & $\q_i \g \to \q_i \gamma$ \\ 114 & $\g \g \to \gamma \gamma$ \\ 115 & $\g \g \to \g \gamma$ \\ \end{tabular} In hadron colliders, processes ISUB = 14 and 29 give the main source of single-$\gamma$ production, with ISUB = 115 giving an additional contribution which, in some kinematics regions, may become important. For $\gamma$-pair production, the process ISUB = 18 is often overshadowed in importance by ISUB = 114. Another source of photons is bremsstrahlung off incoming or outgoing quarks. This has to be treated on an equal footing with QCD parton showering. For time-like parton-shower evolution, i.e. in the final-state showering and in the side branches of the initial-state showering, photon emission may be switched on or off with {\tt MSTJ(41)}. Photon radiation off the space-like incoming quark legs is not yet included, but should be of lesser importance for production at reasonably large $\pT$ values. Radiation off an incoming electron is included in a leading-log approximation. {\bf Warning:} the cross sections for the box graphs 114 and 115 become very complicated, numerically unstable and slow when the full quark mass dependence is included. For quark masses much below the $\hat{s}$ scale, the simplified massless expressions are therefore used --- a fairly accurate approximation. However, there is another set of subtle numerical cancellations between different terms in the massive matrix elements in the region of small-angle scattering. The associated problems have not been sorted out yet. There are therefore two possible solutions. One is to use the massless formulae throughout. The program then becomes faster and numerically stable, but does not give, for example, the characteristic dip (due to destructive interference) at top threshold. This is the current default procedure, with five flavours assumed, but this number can be changed in {\tt MSTP(38)}. The other possibility is to impose cuts on the scattering angle of the hard process, see {\tt CKIN(27)} and {\tt CKIN(28)}, since the numerically unstable regions are when $|\cos\hat{\theta}|$ is close to unity. It is then also necessary to change {\tt MSTP(38)} to 0. \subsubsection{Photoproduction and $\gamma\gamma$ physics} \label{sss:photoprodclass} {\tt MSEL} = 4, 5, 6, 7, 8 \\ ISUB = \\ \begin{tabular}{rl@{\protect\rule{0mm}{\tablinsep}}} 33 & $\q_i \gamma \to \q_i \g$ \\ 34 & $\f_i \gamma \to \f_i \gamma$ \\ 54 & $\g \gamma \to \q_k \qbar_k$ \\ 58 & $\gamma \gamma \to \f_k \fbar_k$ \\ 80 & $\q_i \gamma \to \q_k \pi^{\pm}$ \\ 84 & $\g \gamma \to \Q_k \Qbar_k$ \\ 85 & $\gamma \gamma \to \F_k \Fbar_k$ \\ \end{tabular} An (almost) real photon has both a point-like component and a hadron-like one. This means that several classes of processes may be distinguished, see section \ref{sss:photoprod}. \begin{Enumerate} \item The processes listed above are possible when the photon interacts as a point-like particle, i.e. couples directly to quarks and leptons. \item When the photon acts like a hadron, i.e. is resolved in a partonic substructure, then high-$\pT$ parton--parton interactions are possible, as already described in sections \ref{sss:QCDjetclass} and \ref{sss:promptgammaclass}. These interactions may be further subdivided into VMD and anomalous ones \cite{Sch93,Sch93a}. \item A hadron-like photon can also produce the equivalent of the minimum bias processes of section \ref{sss:minbiasclass}. \end{Enumerate} For $\gamma\p$ events, the best description can be obtained when three separate runs are combined, one for direct, one for VMD and one for anomalous events, see the detailed description in \cite{Sch93,Sch93a}. These correspond to {\tt MSTP(14)} being 0, 2 and 3, respectively. The direct and anomalous components are high-$\pT$ only, while VMD contains both high-$\pT$ and low-$\pT$ events. The option {\tt MSTP(14)=1} combines the VMD and anomalous parts of the photon into one single resolved photon concept, which therefore is less precise than the full subdivision. When combining three runs to obtain the totality of $\gamma\p$ interactions, to the best of our knowledge, it is necessary to choose the $\pT$ cut-offs with some care, so as to represent the expected total cross section. \begin{Itemize} \item The direct processes only depend on the {\tt CKIN(3)} cut-off of the generation, with preferred value 0.5 GeV \cite{Sch93,Sch93a}. Since this value is so low, one must remember to reduce a few other defaults values: {\tt CKIN(1)=2.*CKIN(3)}, {\tt CKIN(5)=CKIN(6)=0.5*CKIN(3)}. For the same reason it is recommended to include a dampening of proton parton distributions, {\tt MSTP(57)=2}. \item The VMD processes work as ordinary hadron--hadron ones, i.e. one obtains both low- and high-$\pT$ events by default, with dividing line set by {\tt PARP(81)} (or {\tt PARP(82)}, depending on minijet unitarization scheme). \item For the anomalous, finally, the minimal $\pT$ of the $\gamma \to \q\qbar$ branching is set in {\tt PARP(15)}. The default is 0.5 GeV, in agreement with the recommended cutoff for the same vertex in direct proccesses. In addition, a lower {CKIN(3)} cut-off should be selected for the hard interactions. This needs some fine-tuning, which in principle should be done separately for each c.m. energy. A good first approximation in the HERA energy range is {\tt CKIN(3)}$=1.50 + 0.0035 \, E_{\mrm{cm}}$. \end{Itemize} The processes in points 1 and 2 can be simulated either with a photon beam or with an electron beam. For a photon beam it is necessary to use option {\tt MSTP(14)} to switch between a point-like and a resolved photon --- it is not possible to simulate the two sets of processes in a single run. An electron by default is assumed to contain photons, but this can be switched off by {\tt MSTP(11)=0}. To have quark and gluon distributions inside the photon (itself inside the electron), {\tt MSTP(12)=1} must be used. For the electron, the two kinds of processes may be generated together, unlike for the photon. It is not possible to have also the low-$\pT$ physics (including multiple interactions in high-$\pT$ events) for an electron beam. Kindly note that subprocess 34 contains both the scattering of an electron off a photon and the scattering of a quark (inside a photon inside an electron) off a photon; the former can be switched off with the help of the {\tt KFIN} array. Currently, hadrons are not defined with any photonic content. None of the processes are therefore relevant in hadron--hadron collisions. In $\e\p$ collisions, the electron can emit an almost real photon, which may interact directly or be resolved. In $\ee$ collisions, one may have direct, singly-resolved or doubly-resolved processes. The low-$\pT$ processes so far have only been implemented for $\gamma \p$ interactions, and not for $\gamma \gamma$ ones. For $\gamma\gamma$, a simple substitute would be to use the $\pi^0$ beam option, in the approximation that a $\pi^0$ is about the same as a $\rho^0$ is about the same as a photon in the vector dominance model. In order to avoid double-counting with the the processes in point 2 above, one should then use {\tt MSTP(82)=2}; additionally it is recommended to put {\tt MSTP(92)=} 2 or 3. Process 54 generates a mixture of quark flavours; allowed flavours are set by the gluon {\tt MDME values}. Process 58 can generate both quark and lepton pairs, according to the {\tt MDME} values of the photon. Processes 84 and 85 are variants of these matrix elements, with fermion masses included in the matrix elements, but where only one flavour can be generated at a time. This flavour is selected as described for processes 81 and 82 in section \ref{sss:heavflavclass}, with the exception that for process 85 the `heaviest' flavour allowed for photon splitting takes to place of the heaviest flavour allowed for gluon splitting. Since lepton KF codes come after quark ones, they are counted as being `heavier', and thus take precedence if they have been allowed. Process 80 is a higher twist one. The theory for such processes is rather shaky, so results should not be taken too literally. The messy formulae given in \cite{Bag82} have no been programmed in full, instead the pion form factor has been parametrized as $Q^2 F_{\pi}(Q^Q) \approx 0.55 / \ln Q^2$, with $Q$ in GeV. \subsubsection{Deep inelastic scattering} \label{sss:DISclass} {\tt MSEL} = 1, 2, 35, 36, 37, 38 \\ ISUB = \\ \begin{tabular}{rl@{\protect\rule{0mm}{\tablinsep}}} 10 & $\f_i \f_j \to \f_k \f_l$ \\ 83 & $\q_i \f_j \to \Q_k \f_l$ \\ \end{tabular} The `deep inelastic scattering' (DIS) processes, i.e. $t$-channel electroweak gauge boson exchange, are traditionally associated with interactions between a lepton or neutrino and a hadron, but processes 10 and 83 can equally well be applied for $\q\q$ scattering in hadron colliders (with a cross section much smaller than corresponding QCD processes, however). If applied to incoming $\ee$ beams, process 10 corresponds to Bhabha scattering. For process 10 both $\gamma$, $\Z^0$ and $\W^{\pm}$ exchange contribute, including interference between $\gamma$ and $\Z^0$. The switch {\tt MSTP(21)} may be used to restrict to only some of these, e.g. neutral or charged current only. In principle, the DIS $x$ variable of an event corresponds to the $x$ value stored in {\tt PARI(33)} or {\tt PARI(34)}, depending on which side the incoming hadron is on, while the DIS $Q^2 = -\hat{t} = ${\tt -PARI(15)}. However, just like initial- and final-state radiation can shift jet momenta, they can modify the momentum of the scattered lepton. Therefore the DIS $x$ and $Q^2$ variables are not automatically conserved. An option, on by default, exists in {\tt MSTP(23)}, where the event can be `modified back' so as to conserve $x$ and $Q^2$, but this option is still rather primitive and should not be taken too literally. Process 83 is the equivalent of process 10 for $\W^{\pm}$ exchange only, but with the heavy-quark mass included in the matrix element. In hadron colliders it is mainly of interest for the production of very heavy flavours, where the possibility of producing just one heavy quark is kinematically favoured over pair production. The selection of the heavy flavour is already discussed in section \ref{sss:heavflavclass}. \subsubsection{Single $\W/\Z$ production} \label{sss:WZclass} {\tt MSEL} = 11, 12, 13, 14, 15, (21) \\ ISUB = \\ \begin{tabular}{rl@{\protect\rule{0mm}{\tablinsep}}} 1 & $\f_i \fbar_i \to \gammaZ$ \\ 2 & $\f_i \fbar_j \to \W^+$ \\ 15 & $\f_i \fbar_i \to \g (\gammaZ)$ \\ 16 & $\f_i \fbar_j \to \g \W^+$ \\ 19 & $\f_i \fbar_i \to \gamma (\gammaZ)$ \\ 20 & $\f_i \fbar_j \to \gamma \W^+$ \\ 30 & $\f_i \g \to \f_i (\gammaZ)$ \\ 31 & $\f_i \g \to \f_k \W^+$ \\ 35 & $\f_i \gamma \to \f_i (\gammaZ)$ \\ 36 & $\f_i \gamma \to \f_k \W^+$ \\ 131 & $\g \g \to \Z^0 \Q_k \Qbar_k$ \\ (141) & $\f_i \fbar_i \to \gamma/\Z^0/\Z'^0$ \\ \end{tabular} This group consists of $2 \to 1$ processes, i.e. production of a single resonance, and $2 \to 2$ processes, where the resonance is recoiling against a jet or a photon. The process 141, which also is listed here, is described further elsewhere. With initial-state showers turned on, the $2 \to 1$ processes also generate additional jets; in order to avoid double-counting, the corresponding $2 \to 2$ processes should therefore not be turned on simultaneously. The basic rule is to use the $2 \to 1$ processes for inclusive generation of $\W/\Z$, i.e. where the bulk of the events studied have $\pT \ll m_{\W/\Z}$, which is where parton showers may be expected to do a good job. For dedicated studies of $\W/\Z$ production at larger transverse momenta, the parton showers tend to underestimate the event rates. It is here better to start from the $2 \to 2$ matrix elements and add showers to these. However, the $2 \to 2$ matrix elements are divergent for $\pT \to 0$, and should not be used down to the low-$\pT$ region, or one may get unphysical cross sections. The problem of double-counting applies not only to $\W/\Z$ production in hadron colliders, but also to a process like $\ee \to \Z^0 \gamma$, which clearly is part of the initial-state radiation corrections to $\ee \to \Z^0$ obtained for {\tt MSTP(11)=1}. As is the case for $\Z$ production in association with jets, the $2 \to 2$ process should therefore only be used for the high-$\pT$ region. The $\Z^0$ of subprocess 1 includes the full interference structure $\gammaZ$; via {\tt MSTP(43)} you can select to produce only $\gamma^*$, only $\Z^0$, or the full $\gammaZ$. The same holds true for the $\Z'^0$ of subprocess 141; via {\tt MSTP(44)} any combination of $\gamma^*$, $\Z^0$ and $\Z'^0$ can be selected. Thus, subprocess 141 with {\tt MSTP(44)=4} is essentially equivalent to subprocess 1 with {\tt MSTP(43)=3}; however, process 141 also includes the possibility of a decay into Higgses. Also processes 15, 19, 30 and 35 contain the full mixture of $\gammaZ$, with {\tt MSTP(43)} available to change this. Only the $\Z^0$ that appears in process 131 does not contain the $\gamma^*$ contribution. Note that process 1, with only $\q\qbar \to \gamma^* \to \ell^+ \ell^-$ allowed, and studied in the region well below the $\Z^0$ mass, is what is conventionally called Drell--Yan. This latter process therefore does not appear under a separate heading, but can be obtained by a suitable setting of switches and parameters. A process like $\f_i \fbar_j \to \gamma \W^+$ is only included in the limit that the $\gamma$ is emitted in the `initial state', while the possibility of a final-state radiation off the $\W^+$ decay products is not explicitly included (but can be obtained implicitly by the parton-shower machinery) and various interference terms are not at all present. Some caution must therefore be exercised; see also section \ref{sss:WZpairclass} for related comments. For the $2 \to 1$ processes, the Breit--Wigner includes an $\hat{s}$-dependent width, which should provide an improved description of line shapes. In fact, from a line-shape point of view, process 1 should provide a more accurate simulation of $\ee$ annihilation events than the dedicated $\ee$ generation scheme of {\Je} (see section \ref{ss:eematrix}). However, the $\pT$ distribution of radiated initial-state photons is probably still better modelled in the {\Je} routines. Another difference is that {\Je} only allows the generation of $\gammaZ \to \q\qbar$, while process 1 additionally contains $\gammaZ \to \ell^+ \ell^-$ and $\gammaZ \to \nu \br{\nu}$. The parton-shower and fragmentation descriptions are the same, but the {\Py} implementation has not been interfaced with the first- and second-order matrix-element options available in {\Je}. Almost all processes in this group have been included with the correct angular distribution in the subsequent $\W/\Z \to \f \fbar$ decays. The exception is process 36, where currently the $\W$ decays isotropically. The process $\ee \to \e^+ \e^- \Z^0$ can be simulated in two different ways. One is to make use of the $\e$ `sea' distribution inside $\e$, i.e. have splittings $\e \to \gamma \to \e$. This can be obtained, together with ordinary $\Z^0$ production, by using subprocess 1, with {\tt MSTP(11)=1} and {\tt MSTP(12)=1}. Then the contribution of the type above is 5.0 pb for a 500 GeV $\ee$ collider, compared with the correct 6.2 pb \cite{Hag91}. Alternatively one may use process 35, with {\tt MSTP(11)=1} and {\tt MSTP(12)=0}. To catch the singularity in the forward direction, regularized by the electron mass, it is necessary to set {\tt CKIN(3)=CKIN(5)=0.01} --- using lower values will only slow down execution, not significantly increasing the cross section. One then obtains 5.1 pb, i.e. again 20\% below the correct value, but now also generates a $\pT$ distribution for the $\Z^0$; this is therefore to be preferred. Process 36, $\f \gamma \to \f' \W^{\pm}$ may have corresponding problems; except that in $\ee$ the forward scattering amplitude for $\e \gamma \to \nu \W$ is killed (radiation zero), which means that the differential cross section is vanishing for $\pT \to 0$. It is therefore feasible to use the default {\tt CKIN(3)} and {\tt CKIN(5)} values in $\ee$, and one also comes closer to the correct cross section. One single true $2 \to 3$ process is included in this class as well; namely $\g \g \to \Z^0 \Q \Qbar$, with full massive matrix elements. The more complicated phase space and the lengthy matrix-element evaluations make this process extremely slow. With the quark flavour picked to be $\b$, it may form an important background to intermediate mass Higgs searches in the multilepton channel. The quark flavour is stored in {\tt KFPR(131,2)}; the default is $5 = \b$. The kinematics is set up in terms of a $\Z^0$ recoiling against the $\Q \Qbar$ system, and all ordinary kinematics cut for a $2 \to 2$ process can be used on this level, including {\tt CKIN(43)} and {\tt CKIN(44)} to restrict the range of the $\Q \Qbar$ invariant mass. In addition, for this process alone, {\tt CKIN(51) - CKIN(54)} can be used to set the $\pT$ range of the two quarks; as is to be expected, that of the $\Z^0$ is set by {\tt CKIN(3) - CKIN(4)}. Since the optimization procedure is not set up to probe the full multidimensional phase space allowed in this process, maximum violations may be quite large. It may then be useful to make a preliminary run to find how big the violations are in total, and then use the {\tt MSTP(121)=1} option in the full run. \subsubsection{$\W/\Z$ pair production} \label{sss:WZpairclass} {\tt MSEL} = 15 \\ ISUB = \\ \begin{tabular}{rl@{\protect\rule{0mm}{\tablinsep}}} 22 & $\f_i \fbar_i \to (\gammaZ) (\gammaZ)$ \\ 23 & $\f_i \fbar_j \to \Z^0 \W^+$ \\ 25 & $\f_i \fbar_i \to \W^+ \W^-$ \\ 69 & $\gamma \gamma \to \W^+ \W^-$ \\ 70 & $\gamma \W^+ \to \Z^0 \W^+$ \\ \end{tabular} In this section we mainly consider the production of $\W/\Z$ pairs by fermion--antifermion annihilation, but also include two processes which involve $\gamma/\W$ beams. Scatterings between gauge-boson pairs, i.e. processes like $\W^+ \W^- \to \Z^0 \Z^0$, depend so crucially on the assumed Higgs scenario that they are considered separately in section \ref{sss:heavySMHclass}. The cross sections used for the above processes are those derived in the narrow-width limit, but have been extended to include Breit--Wigner shapes with mass-dependent widths. However, one should realize that other graphs, not included here, can contribute in regions away from the $\W/\Z$ mass. This problem is especially important if several flavours coincide in the four-fermion final state. Consider, as an example, $\ee \to \mu^+ \mu^- \nu_{\mu} \br{\nu}_{\mu}$. Not only would such a final state receive contributions from intermediate $\Z^0\Z^0$ and $\W^+\W^-$ states, but also from processes $\ee \to \Z^0 \to \mu^+ \mu^-$, followed either by $\mu^+ \to \mu^+ \Z^0 \to \mu^+ \nu_{\mu} \br{\nu}_{\mu}$, or by $\mu^+ \to \br{\nu}_{\mu} \W^+ \to \br{\nu}_{\mu} \mu^+ \nu_{\mu}$. In addition, all possible interferences should be considered. Since this is not done, the processes have to be used with some sound judgement. Very often, one may wish to constrain a lepton pair mass to be close to $m_{\Z}$, in which case a number of the possible `other' processes are negligible. Of the above processes, the first contains the full $\f_i \fbar_i \to (\gammaZ)(\gammaZ)$ structure, obtained by a straightforward generalization of the formulae in ref. \cite{Gun86} (done by the present author). Of course, the possibility of there being significant contributions from graphs that are not included is increased, in particular if one $\gamma^*$ is very light and therefore could be a bremsstrahlung-type photon. It is possible to use {\tt MSTP(43)} to recover the pure $\Z^0$ case, i.e. $\f_i \fbar_i \to \Z^0 \Z^0$ exclusively. In processes 23 and 70, only the pure $\Z^0$ contribution is included. Full angular correlations are included for the first three processes, i.e. the full $2 \to 2 \to 4$ matrix elements are included in the resonance decays, including the appropriate $\gammaZ$ interference in process 22. In the latter two processes no spin information is currently preserved, i.e. the $\W/\Z$ bosons are allowed to decay isotropically. We remind you that the mass ranges of the two resonances may be set with the {\tt CKIN(41) - CKIN(44)} parameters; this is particularly convenient, for instance, to pick one resonance almost on the mass shell and the other not. \subsection{Higgs Production} A fair fraction of all the processes in {\Py} deal with Higgs production in one form or another. This multiplication is caused by the need to consider production by several different processes, depending on Higgs mass and machine type. Further, the program contains a full two-Higgs-multiplet scenario, as predicted for example in the Minimal Supersymmetric extension of the Standard Model (MSSM). Therefore the continued discussion is, somewhat arbitrarily, subdivided into a few different scenarios. \subsubsection{Light Standard Model Higgs} \label{sss:lightSMHclass} {\tt MSEL} = 16, 17, 18 \\ ISUB = \\ \begin{tabular}{rl@{\protect\rule{0mm}{\tablinsep}}} 3 & $\f_i \fbar_i \to \H^0$ \\ 24 & $\f_i \fbar_i \to \Z^0 \H^0$ \\ 26 & $\f_i \fbar_j \to \W^+ \H^0$ \\ 102 & $\g \g \to \H^0$ \\ 103 & $\gamma \gamma \to \H^0$ \\ 110 & $\f_i \fbar_i \to \gamma \H^0$ \\ 111 & $\f_i \fbar_i \to \g \H^0$ \\ 112 & $\f_i \g \to \f_i \H^0$ \\ 113 & $\g \g \to \g \H^0$ \\ 121 & $\g \g \to \Q_k \Qbar_k \H^0$ \\ 122 & $\q_i \qbar_i \to \Q_k \Qbar_k \H^0$ \\ 123 & $\f_i \f_j \to \f_i \f_j \H^0$ ($\Z^0 \Z^0$ fusion) \\ 124 & $\f_i \f_j \to \f_k \f_l \H^0$ ($\W^+ \W^-$ fusion) \\ \end{tabular} In this section we discuss the production of a reasonably light Standard Model Higgs, below 700 GeV, say, so that the narrow width approximation can be used with some confidence. Below 400 GeV there would certainly be no trouble, while above that the narrow width approximation is gradually starting to break down. In a hadron collider, the main production processes are 102, 123 and 124, i.e. $\g\g$, $\Z^0 \Z^0$ and $\W^+ \W^-$ fusion. In the latter two processes, it is also necessary to take into account the emission of the space-like $\W/\Z$ bosons off quarks, which in total gives the $2 \to 3$ processes above. Further processes of lower cross sections may be of interest because of easier signals. For instance, processes 24 and 26 give associated production of a $\Z$ or a $\W$ together with the $\H^0$. There is also the processes 3, 121 and 122, which involve production of heavy flavours. Process 3 contains contributions from all flavours, but is completely dominated by the subprocess $\t \tbar \to \H^0$, i.e. by the contribution from the top sea distributions. This process is by now known to overestimate the cross section for Higgs production as compared with a more careful calculation based on the subprocess $\g \g \to \t \tbar \H^0$ (121). The difference between the two is that in process 3 the $\t$ and $\tbar$ are added by the initial-state shower, while in 121 the full matrix element is used. The price to be paid is that the complicated multibody phase space in process 121 makes the program run slower than with most other processes. One should therefore think twice before using it. As usual, it would be double-counting to include both 3 and 121. Process 122 is similar in structure to 121, but is less important. A subprocess like 113, with a Higgs recoiling against a gluon jet, is also effectively generated by initial-state corrections to subprocess 102; thus, in order to avoid double-counting, just as for the case of $\Z^0/\W^+$ production, section \ref{sss:WZclass}, these subprocesses should not be switched on simultaneously. Process 102 should be used for inclusive production of Higgs, and 111--113 for the study of the Higgs subsample with high transverse momentum. In $\ee$ annihilation, associated production of an $\H^0$ with a $\Z^0$, process 24, is usually the dominant one close to threshold, while the $\Z^0 \Z^0$ and $\W^+ \W^-$ fusion processes 123 and 124 win out at high energies. Process 103, $\gamma\gamma$ fusion, may also be of interest, in particular when the possibilities of beamstrahlung photons and backscattered photons are included. Process 110, which gives an $\H^0$ in association with a $\gamma$, is a loop process and is therefore suppressed in rate. Only for a rather massive $\H^0$ (mass above 60 GeV at LEP 1) can it start to compete with the associated production of a $\Z^0$, since phase space suppression is less severe for the former than for the latter. The branching ratios of the Higgs are very strongly dependent on the mass. In principle, the program is set up to calculate these correctly, as a function of the actual Higgs mass, i.e. not just at the nominal mass. However, higher-order corrections may at times be important and not fully unambiguous; see for instance {\tt MSTP(37)}. Since the Higgs is a spin-0 particle it decays isotropically. In decay processes such as $\H^0 \to \W^+ \W^- \to 4$ fermions angular correlations are included. Also in processes 24 and 26, $\Z^0$ and $\W^{\pm}$ decay angular distributions are correctly taken into account. \subsubsection{Heavy Standard Model Higgs} \label{sss:heavySMHclass} ISUB = \\ \begin{tabular}{rl@{\protect\rule{0mm}{\tablinsep}}} 5 & $\Z^0 \Z^0 \to \H^0$ \\ 8 & $\W^+ \W^- \to \H^0$ \\ 71 & $\Z^0 \Z^0 \to \Z^0 \Z^0$ (longitudinal) \\ 72 & $\Z^0 \Z^0 \to \W^+ \W^-$ (longitudinal) \\ 73 & $\Z^0 \W^+ \to \Z^0 \W^+$ (longitudinal) \\ 76 & $\W^+ \W^- \to \Z^0 \Z^0$ (longitudinal) \\ 77 & $\W^+ \W^{\pm} \to \W^+ \W^{\pm}$ (longitudinal) \\ \end{tabular} Processes 5 and 8 are the simple $2 \to 1$ versions of what is now available in 123 and 124 with the full $2 \to 3$ kinematics. For low Higgs masses processes 5 and 8 overestimate the correct cross sections and should not be used, whereas good agreement between the $2 \to 1$ and $2 \to 3$ descriptions is observed when heavy Higgs production is studied. The subprocesses 5 and 8, $V V \to \H^0$, which contribute to the processes $V V \to V' V'$, show a bad high-energy behaviour. Here $V$ denotes a longitudinal intermediate gauge boson, $\Z^0$ or $\W^{\pm}$. This can be cured only by the inclusion of all $V V \to V' V'$ graphs, as is done in subprocesses 71, 72, 73, 76 and 77. In particular, subprocesses 5 and 8 give rise to a fictitious high-mass tail of the Higgs. If this tail is thrown away, however, the agreement between the $s$-channel graphs only (subprocesses 5 and 8) and the full set of graphs (subprocesses 71 etc.) is very good: for a Higgs of nominal mass 300 (800) GeV, a cut at 600 (1200) GeV retains 95\% (84\%) of the total cross section, and differs from the exact calculation, cut at the same values, by only 2\% (11\%) (numbers for SSC energies). With this prescription there is therefore no need to use subprocesses 71 etc. rather than subprocesses 5 and 8. For subprocess 77, there is an option, see {\tt MSTP(45)}, to select the charge combination of the scattering $\W$'s: like-sign, opposite-sign (relevant for Higgs), or both. Process 77 contains a divergence for $\pT \to 0$ due to $\gamma$-exchange contributions. This leads to an infinite total cross section, which is entirely fictitious, since the simple parton-distribution function approach to the longitudinal $\W$ flux is not appropriate in this limit. For this process, it is therefore necessary to make use of a cut, e.g. $\pT > m_{\W}$. For subprocesses 71, 72, 76 and 77, an option is included (see {\tt MSTP(46)}) whereby the user can select only the $s$-channel Higgs graph; this will then be essentially equivalent to running subprocess 5 or 8 with the proper decay channels (i.e. $\Z^0\Z^0$ or $\W^+\W^-$) set via {\tt MDME}. The difference is that the Breit--Wigners in subprocesses 5 and 8 contain a mass-dependent width, whereas the width in subprocesses 71--77 is calculated at the nominal Higgs mass; also, higher-order corrections to the widths are treated more accurately in subprocesses 5 and 8. Further, processes 71--77 assume the incoming $\W/\Z$ to be on the mass shell, with associated kinematics factors, while processes 5 and 8 have $\W/\Z$ correctly space-like. All this leads to differences in the cross sections by up to a factor of 1.5. In the absence of a Higgs, the sector of longitudinal $\Z$ and $\W$ scattering will become strongly interacting at energies above 1 TeV. The models proposed by Dobado, Herrero and Terron \cite{Dob91} to describe this kind of physics have been included as alternative matrix elements for subprocesses 71, 72, 73, 76 and 77, selectable by {\tt MSTP(46)}. From the point of view of the general classification scheme for subprocesses, this kind of models should appropriately be included as separate subprocesses with numbers above 100, but the current solution allows a more efficient reuse of existing code. By a proper choice of parameters, it is also here possible to simulate the production of a techni-$\rho$. Currently, the scattering of transverse gauge bosons has not been included, neither that of mixed transverse--longitudinal scatterings. These are expected to be less important at high energies, and do not contain an $\H^0$ resonance peak, but need not be entirely negligible in magnitude. As a rule of thumb, processes 71--77 should not be used for $VV$ invariant masses below 500 GeV. The decay products of the longitudinal gauge bosons are correctly distributed in angle. \subsubsection{Extended neutral Higgs sector} \label{sss:extneutHclass} {\tt MSEL} = 19 \\ ISUB = \\ \begin{tabular}{rrrl@{\protect\rule{0mm}{\tablinsep}}} $\H^0$ & $\H'^0$ & $\A^0$ & \\ 3 & 151 & 156 & $\f_i \fbar_i \to X$ \\ 102 & 152 & 157 & $\g \g \to X$ \\ 103 & 153 & 158 & $\gamma \gamma \to X$ \\ 24 & 171 & 176 & $\f_i \fbar_i \to \Z^0 X$ \\ 26 & 172 & 177 & $\f_i \fbar_j \to \W^+ X$ \\ 123 & 173 & 178 & $\f_i \f_j \to \f_i \f_j X$ ($\Z \Z$ fusion) \\ 124 & 174 & 179 & $\f_i \f_j \to \f_k \f_l X$ ($\W^+\W^-$ fusion) \\ 121 & 181 & 186 & $\g \g \to \Q_k \Qbar_k X$ \\ 122 & 182 & 187 & $\q_i \qbar_i \to \Q_k \Qbar_k X$ \\ \end{tabular} \\ ISUB = \\ \begin{tabular}{rl@{\protect\rule{0mm}{\tablinsep}}} (141) & $\f_i \fbar_i \to \gamma/\Z^0/\Z'^0$ \\ \end{tabular} In {\Py}, the particle content of a two-Higgs-doublet scenario is included: two neutral scalar particles, 25 and 35, one pseudoscalar one, 36, and a charged doublet, $\pm 37$. (Of course, these particles may also be associated with corresponding Higgs states in larger multiplets.) By convention, we choose to call the lighter scalar Higgs $\H^0$ and the heavier $\H'^0$ --- this differs from the convention in the MSSM, where the lighter is called $\hrm^0$ and the heavier $\H^0$, but allows us to call the Higgs of the one-Higgs scenario $\H^0$. The pseudoscalar is called $\A^0$ and the charged $\H^{\pm}$. Charged-Higgs production is covered in section \ref{sss:chHclass}. A number of $\H^0$ processes have been duplicated for $\H'^0$ and $\A^0$. The correspondence between ISUB numbers is shown in the table above: the first column of ISUB numbers corresponds to $X = \H^0$, the second to $X = \H'^0$, and the third to $X = \A^0$. Note that several of these processes are not expected to take place at all, owing to vanishing Born term couplings. We have still included them for flexibility in simulating arbitrary couplings at the Born or loop level. A few Standard Model Higgs processes have no correspondence in the scheme above. These include \begin{Itemize} \item 5 and 8, which anyway have been superseded by 123 and 124; \item 71, 72, 73, 76 and 77, which deal with what happens if there is no light Higgs, and so is a scenario complementary to the one above, where several light Higgses are assumed; \item 110, which is mainly of interest in Standard Model Higgs searches; and \item 111, 112 and 113, which describe the high-$\pT$ tail of the Higgs production, and are less interesting for most Higgs studies. \end{Itemize} In processes 121, 122, 181, 182, 186 and 187 the recoiling heavy flavour is assumed to be top, which is the only one of interest in the Standard Model, and the one where the parton-distribution-function approach invoked in processes 3, 151 and 156 is least reliable. However, it is possible to change the quark flavour in 121 etc.; for each process ISUB this flavour is given by {\tt KFPR(ISUB,2)}. This may become relevant if couplings to $\b\bbar$ states are enhanced, e.g. if $\tan\beta \gg 1$ in the MSSM. By default, the $\H^0$ has the couplings of the Standard Model Higgs, while the $\H'^0$ and $\A^0$ have couplings set in {\tt PARU(171) - PARU(178)} and {\tt PARU(181) - PARU(190)}, respectively. The default values for the $\H'^0$ and $\A^0$ have no deep physics motivation, but are set just so that the program will not crash due to the absence of any couplings whatsoever. You should therefore set the above couplings to your desired values if you want to simulate either $\H'^0$ or $\A^0$. Also the couplings of the $\H^0$ particle can be modified, in {\tt PARU(161) - PARU(165)}, provided that {\tt MSTP(4)} is set to 1. For {\tt MSTP(4)=2}, the mass of the $\H^0$ (in {\tt PMAS(25,1)}) and the $\tan\beta$ value (in {\tt PARU(141)}) are used to derive the masses of the other Higgses, as well as all Higgs couplings. {\tt PMAS(35,1) - PMAS(37,1)} and {\tt PARU(161) - PARU(195)} are overwritten accordingly. The relations used are the ones of the Born-level MSSM \cite{Gun90}. Today, loop corrections to those expressions have been calculated, and are known to have non-negligible effects on the resulting phenomenology. Eventually the modified relations will be included as an additional option, but this has not yet been done. Note that not all combinations of $m_{\H}$ and $\tan\beta$ are allowed; the requirement of a finite $\A^0$ mass imposes the constraint \begin{equation} m_{\H} < m_{\Z} \, \frac{\tan^2\beta - 1}{\tan^2\beta + 1}, \end{equation} or, equivalently, \begin{equation} \tan^2\beta > \frac{m_{\Z} + m_{\H}}{m_{\Z} - m_{\H}}. \end{equation} If this condition is not fulfilled, the program will crash. Process 141 can also be used to simulate $\Z^0 \to \H^0 \A^0$ and $\Z^0 \to \H'^0 \A^0$ for associated neutral Higgs production. The fact that we here make use of the $\Z'^0$ can easily be discounted, either by letting the relevant couplings vanish, or by the option {\tt MSTP(44)=4}. Finally, heavier Higgses may decay into lighter ones, if kinematically allowed, in processes like $\A^0 \to \Z^0 \H^0$ or $\H^+ \to \W^+ \H^0$. Such modes are included as part of the general mixture of decay channels, but they can be enhanced if the uninteresting channels are switched off. \subsubsection{Charged Higgs sector} \label{sss:chHclass} {\tt MSEL} = 23 \\ ISUB = \\ \begin{tabular}{rl@{\protect\rule{0mm}{\tablinsep}}} 143 & $\f_i \fbar_j \to \H^+$ \\ 161 & $\f_i \g \to \f_k \H^+$ \\ (141) & $\f_i \fbar_i \to \gamma/\Z^0/\Z'^0$ \\ \end{tabular} A charged Higgs doublet, $\H^{\pm}$, is included in the program. This doublet may be the one predicted in the MSSM scenario, see section \ref{sss:extneutHclass}, or in any other scenario. The $\tan\beta$ parameter, which is relevant also for charged Higgs couplings, is set via {\tt PARU(141)}. The basic subprocess for charged Higgs production in hadron colliders is ISUB = 143. However, this process is dominated by $\t \bbar \to \H^+$, and so depends on the choice of $\t$ parton distribution. A better representation is provided by subprocess 161, $\f \g \to \f' \H^+$; i.e. actually $\bbar \g \to \tbar \H^+$. It is therefore recommended to use 161 and not 143; to use both would be double-counting. In subprocess 141, the decay $\gamma^*/\Z^0/\Z'^0 \to \H^+ \H^-$ allows the production of a pair of charged Higgs particles. This process is especially important in $\ee$ colliders. The coupling of the $\gamma^*$ to $\H^+ \H^-$ is determined by the charge alone, while the $\Z^0$ coupling is regulated by {\tt PARU(142)}, and that of the $\Z'^0$ by {\tt PARU(143)}. The $\Z'^0$ piece can be switched off, e.g. by {\tt MSTP(44)=4}. An ordinary $\Z^0$, i.e. particle code 23, cannot be made to decay into $\H^+ \H^-$, however. A major potential source of charged Higgs production is top decay. When the top is treated as a resonance (the default option), it is possible to switch on the decay channel $\t \to \b \H^+$. Top will then decay to $\H^+$ a fraction of the time, whichever way it is produced. The branching ratio is automatically calculated, based on the $\tan\beta$ value and masses. It is possible to only have the $\H^+$ decay mode switched on, in which case the cross section is reduced accordingly. If one instead assumes that top hadrons are formed, branching ratios are not automatically calculated. However, you can set, for the generic top hadron 86, the branching ratios for the two main channels $\t \to \b \H^+$ and $\t \to \b \W^+$. In this option the cross section for top production will not be reduced if only the $\t \to \b \H^+$ decay is switched on, cf. section \ref{sss:resdecaycross}. \subsection{Non-Standard Physics} The number of possible non-Standard Model scenarios is essentially infinite, but many of the studied scenarios still share a lot of aspects. For instance, new $\W'$ and $\Z'$ gauge bosons can arise in a number of different ways. Therefore it still makes sense to try to cover a few basic classes of particles, with enough freedom in couplings that many kinds of detailed scenarios can be accommodated by suitable parameter choices. We have already seen one example of this, in the extended Higgs sector above. In this section a few other kinds of non-standard generic physics is discussed. Clearly many others could have been included, but there is probably only one glaring omission: currently no supersymmetric particle production has been included. One main reason for this is the large number of particles, processes, possible mass hierarchies and decay chains. \subsubsection{Fourth-generation fermions} {\tt MSEL} = 7, 8, 37, 38 \\ ISUB = \\ \begin{tabular}{rl@{\protect\rule{0mm}{\tablinsep}}} 1 & $\f_i \fbar_i \to \gammaZ$ \\ 2 & $\f_i \fbar_j \to \W^+$ \\ 81 & $\q_i \qbar_i \to \Q_k \Qbar_k$ \\ 82 & $\g \g \to \Q_k \Qbar_k$ \\ 83 & $\q_i \f_j \to \Q_k \f_l$ \\ 84 & $\g \gamma \to \Q_k \Qbar_k$ \\ 85 & $\gamma \gamma \to \F_k \Fbar_k$ \\ 141 & $\f_i \fbar_i \to \gamma/\Z^0/\Z'^0$ \\ 142 & $\f_i \fbar_j \to \W'^+$ \\ \end{tabular} The prospects of a fourth generation currently seem rather dim, but the appropriate flavour content is still found in the program. In fact, the fourth generation is included on an equal basis with the first three, provided {\tt MSTP(1)=4}. Also processes other than the ones above can therefore be used, e.g. all other processes with gauge bosons, including non-standard ones such as the $\Z'^0$. We therefore do not repeat the descriptions found elsewhere, e.g. how to set only the desired flavour in processes 81--85. Note that it may be convenient to set {\tt CKIN(1)} and other cuts such that the mass of produced gauge bosons is enough for the wanted particle production --- in principle the program will cope even without that, but possibly at the expense of very slow execution. \subsubsection{New gauge bosons} {\tt MSEL} = 21, 22, 24 \\ ISUB = \\ \begin{tabular}{rl@{\protect\rule{0mm}{\tablinsep}}} 141 & $\f_i \fbar_i \to \gamma/\Z^0/\Z'^0$ \\ 142 & $\f_i \fbar_j \to \W'^+$ \\ 144 & $\f_i \fbar_j \to \R$ \\ \end{tabular} The $\Z'^0$ of subprocess 141 contains the full $\gamma^*/\Z^0/\Z'^0$ interference structure for couplings to fermion pairs. With {\tt MSTP(44)} it is possible to pick only a subset, e.g. only the pure $\Z'^0$ piece. The couplings of the $\Z'^0$ to quarks and leptons can be set via {\tt PARU(121) - PARU(128)}. The eight numbers correspond to the vector and axial couplings of down-type quarks, up-type quarks, leptons and neutrinos, respectively. The default corresponds to the same couplings as that of the Standard Model $\Z^0$, with axial couplings $a_{\f} = \pm 1$ and vector couplings $v_{\f} = a_{\f} - 4 e_{\f} \ssintw$. This implies a resonance width that increases linearly with the mass. By a suitable choice of the parameters, it is possible to simulate just about any imaginable $\Z'^0$ scenario, with full interference effects in cross sections and decay angular distributions. The coupling to the decay channel $\Z'^0 \to \W^+ \W^-$ is regulated by {\tt PARU(129) - PARU(130)}. The former gives the strength of the coupling, which determines the rate. The default, {\tt PARU(129)=1.}, corresponds to the `extended gauge model' of \cite{Alt89}, wherein the $\Z^0 \to \W^+ \W^-$ coupling is used, scaled down by a factor $m_{\W}^2/m_{\Z'}^2$, to give a $\Z'^0$ partial width into this channel that again increases linearly. If this factor is cancelled, by having {\tt PARU(129)} proportional to $m_{\Z'}^2/m_{\W}^2$, one obtains a partial width that goes like the fifth power of the $\Z'^0$ mass, the `reference model' of \cite{Alt89}. In the decay angular distribution one could imagine a much richer structure than is given by the one parameter {\tt PARU(130)}. Other decay modes include $\Z'^0 \to \Z^0 \H^0$, predicted in left--right symmetric models (see {\tt PARU(145)} and ref. \cite{Coc91}), and a number of other Higgs decay channels, see sections \ref{sss:extneutHclass} and \ref{sss:chHclass}. The $\W'^{\pm}$ of subprocess 142 so far does not contain interference with the Standard Model $\W^{\pm}$ --- in practice this should not be a major limitation. The couplings of the $\W'$ to quarks and leptons are set via {\tt PARU(131) - PARU(134)}. Again one may set vector and axial couplings freely, separately for the $\q \qbar'$ and the $\ell \nu_{\ell}$ decay channels. The defaults correspond to the $V-A$ structure of the Standard Model $\W$, but can be changed to simulate a wide selection of models. One possible limitation is that the same Cabibbo--Kobayashi--Maskawa quark mixing matrix is assumed as for the standard $\W$. The coupling $\W' \to \Z^0 \W$ can be set via {\tt PARU(135) - PARU(136)}. Further comments on this channel as for $\Z'$; in particular, default couplings again agree with the `extended gauge model' of \cite{Alt89}. A $\W' \to \W \H^0$ channel is also included, in analogy with the $\Z'^0 \to \Z^0 \H^0$ one, see {\tt PARU(146)}. The $\R$ boson (particle code 40) of subprocess 144 represents one possible scenario for a horizontal gauge boson, i.e. a gauge boson that couples between the generations, inducing processes like $\s \dbar \to \R^0 \to \mu^- \e^+$. Experimental limits on flavour-changing neutral currents forces such a boson to be fairly heavy. The model implemented is the one described in \cite{Ben85a}. \subsubsection{Leptoquarks} \label{sss:LQclass} {\tt MSEL} = 25 \\ ISUB = \\ \begin{tabular}{rl@{\protect\rule{0mm}{\tablinsep}}} 145 & $\q_i \ell_j \to \L_{\Q}$ \\ 162 & $\q \g \to \ell \L_{\Q}$ \\ 163 & $\g \g \to \L_{\Q} \br{\L}_{\Q}$ \\ 164 & $\q_i \qbar_i \to \L_{\Q} \br{\L}_{\Q}$ \\ \end{tabular} Several processes that can generate a leptoquark have been included. Currently only one leptoquark has been implemented, as particle 39, denoted $\L_{\Q}$. The leptoquark is assumed to carry specific quark and lepton quantum numbers, by default $\u$ quark plus electron. These flavour numbers are conserved, i.e. a process such as $\u \e^- \to \L_{\Q} \to \d \nu_{\e}$ is not allowed. This may be a bit restrictive, but it represents one of many leptoquark possibilities. The spin of the leptoquark is assumed to be zero, i.e. its decay is isotropical. Although only one leptoquark is implemented, its flavours may be changed arbitrarily to study the different possibilities. The flavours of the leptoquark are defined by the quark and lepton flavours in the decay mode list. Since only one decay channel is allowed, this means that the quark flavour is stored in {\tt KFDP(MDCY(39,2),1)} and the lepton one in {\tt KFDP(MDCY(39,2),2)}. The former must always be a quark, while the latter could be a lepton or an antilepton; a charge-conjugate partner is automatically defined by the program. At initialization, the charge is recalculated as a function of the flavours defined; also the leptoquark name is redefined to be of the type {\tt 'LQ\_(q)(l)'}, where actual quark {\tt (q)} and lepton {\tt (l)} flavours are displayed. The $\L_{\Q} \to \q \ell$ vertex contains an undetermined Yukawa coupling strength, which affects both the width of the leptoquark and the cross section for many of the production graphs. This strength may be changed in {\tt PARU(151)}. The definition of {\tt PARU(151)} corresponds to the $k$ factor of \cite{Hew88}, i.e. to $\lambda^2/(4\pi\alphaem)$, where $\lambda$ is the Yukawa coupling strength of \cite{Wud86}. Note that {\tt PARU(151)} is thus quadratic in the coupling. The leptoquark is likely to be fairly long-lived, in which case it has time to fragment into a mesonic- or baryonic-type state, which would decay later on. This is a bit tedious to handle; therefore the leptoquark is always assumed to decay before fragmentation has to be considered. This may give some imperfections in the event generation, but should not be off by much in the final analysis. Inside the program, the leptoquark is treated as a resonance. Since it carries colour, some extra care is required. In particular, it is not allowed to put the leptoquark stable, by modifying either {\tt MDCY(39,1)} or {\tt MSTP(41)}: then the leptoquark would be handed undecayed to {\Je}, which would try to fragment it (as it does with any other coloured object), and most likely crash. \subsubsection{Compositeness and anomalous couplings} \label{sss:ancoupclass} ISUB = \\ \begin{tabular}{rl@{\protect\rule{0mm}{\tablinsep}}} 11 & $\f_i \f_j \to \f_i \f_j$ (QCD) \\ 12 & $\f_i \fbar_i \to \f_k \fbar_k$ \\ 20 & $\f_i \fbar_j \to \gamma \W^+$ \\ 165 & $\f_i \fbar_i \to \f_k \fbar_k$ (via $\gammaZ$) \\ 166 & $\f_i \fbar_j \to \f_k \fbar_l$ (via $\W^{\pm}$) \\ \end{tabular} Some processes have been set up to allow anomalous coupling to be introduced, in addition to the Standard Model ones. These can be switched on by {\tt MSTP(5)}$\geq 1$; the default {\tt MSTP(5)=0} corresponds to the Standard Model behaviour. In processes 11 and 12, the quark substructure is included in the left--left isoscalar model \cite{Eic84} for {\tt MSTP(5)=1}, with compositeness scale $\Lambda$ given in {\tt PARU(155)} (default 1000~GeV) and sign $\eta$ of interference term in {\tt PARU(156)} (default $+1$; only other alternative $-1$). The above model assumes that only $\u$ and $\d$ quarks are composite (at least at the scale studied); with {\tt MSTP(5)=2} compositeness terms are included in the interactions between all quarks. The processes 165 and 166 are basically equivalent to 1 and 2, i.e. $\gammaZ$ and $\W^{\pm}$ exchange, respectively, but a bit less fancy (no mass-dependent width etc.). The reason for this duplication is that the resonance treatment formalism of processes 1 and 2 could not easily be extended to include other than $s$-channel graphs. In processes 165 and 166, only one final-state flavour is generated at the time; this flavour should be set in {\tt KFPR(165,1)} and {\tt KFPR(166,1)}, respectively. For process 166 one gives the down-type flavour, and the program will associate the up-type flavour of the same generation. Defaults are 11 in both cases, i.e. $\ee$ and $\e^+ \nu_{\e}$ ($\e^- \br{\nu}_{\e}$) final states. While {\tt MSTP(5)=0} gives the Standard Model results, {\tt MSTP(5)=1} contains the left--left isoscalar model (which does not affect process 166), and {\tt MSTP(5)=3} the helicity-non-conserving model (which affects both) \cite{Eic84,Lan91}. Both models above assume that only $\u$ and $\d$ quarks are composite; with {\tt MSTP(5)=} 2 or 4, respectively, contact terms are included for all quarks in the initial state. Parameters are {\tt PARU(155)} and {\tt PARU(156)}, as above. Note that processes 165 and 166 are bookkept as $2 \to 2$ processes, while 1 and 2 are $2 \to 1$ ones. This means that the default $\Q^2$ scale in parton distributions is $\pT^2$ for the former and $\hat{s}$ for the latter. To make contact between the two, it is recommended to set {\tt MSTP(32)=4}, so as to use $\hat{s}$ as scale also for processes 165 and 166. In process 20, for $\W \gamma$ pair production, it is possible to set an anomalous magnetic moment for the $\W$ in {\tt PARU(153)} ($= \eta = \kappa-1$; where $\kappa = 1$ is the Standard Model value). The production process is affected according to the formulae of \cite{Sam91}, while $\W$ decay currently remains unaffected. It is necessary to set {\tt MSTP(5)=1} to enable this extension. \subsubsection{Excited fermions} \label{sss:qlstarclass} ISUB = \\ \begin{tabular}{rl@{\protect\rule{0mm}{\tablinsep}}} 147 & $\d \g \to \d^*$ \\ 148 & $\u \g \to \u^*$ \\ \end{tabular} Compositeness scenarios may also give rise to sharp resonances of excited quarks and leptons. If {\tt MSTP(6)=1}, then at initialization the standard fourth generation of fermions will be overwritten, and made to correspond to an excited copy of the first generation, consisting of spin $1/2$ particles $\d^*$ (code 7), $\u^*$ (8), $\e^*$ (17) and $\nu^*_{\e}$ (18). Since the original fourth-generation information is lost, it is then not possible to generate fourth-generation particles in the same run. The current implementation only contains gauge interactions, i.e. no contact interactions. The couplings $f$, $f'$ and $f_s$ to the {\bf SU(2)}, {\bf U(1)} and {\bf SU(3)} groups are stored in {\tt PARU(157) - PARU(159)}, the scale parameter $\Lambda$ in {\tt PARU(155)}; you are also expected to change the $\f^*$ masses in accordance with what is desired --- see \cite{Bau90} for details on conventions. Decay processes are of the types $\q^* \to \q \g$, $\q^* \to \q \gamma$, $\q^* \to \q \Z^0$ or $\q^* \to \q' \W^{\pm}$. Production is currently only by quark-gluon fusion. \subsubsection{Technicolor} \label{sss:technicolorclass} ISUB = \\ \begin{tabular}{rl@{\protect\rule{0mm}{\tablinsep}}} 149 & $\g \g \to \eta_{\mrm{techni}}$ \\ \end{tabular} The technicolor scenario offers an alternative to the ordinary Higgs mechanism for giving masses to the $\W$ and $\Z$. The technicolor gauge group is an analogue of QCD, with a rich spectrum of technimesons made out of techniquarks. Three of the technipions assume the role of the longitudinal components of the $\W$ and $\Z$ bosons, but many other states remain as separate particles. No fully realistic model has been found so far, however, so any phenomenology has to be taken as indicative only. In section \ref{sss:heavySMHclass} it is discussed how processes 71--77, in some of its options, can be used to simulate a scenario with techni-$\rho$ resonances in longitudinal gauge boson scattering. Here we present another process, that of the production of a techni-$\eta$. This particle has zero spin, is a singlet under electroweak {\bf SU(2)}$\times${\bf U(1)}, but carries octet colour charge. It is one of the possible techni-$\pi$ particles; the name techni-$\eta$ is part of a subclassification not used by all authors. The techni-$\eta$ couples to ordinary fermions according to the fermion squared mass. The dominant decay mode is therefore $\t \tbar$, if allowed. The coupling to a $\g \g$ state is roughly comparable with that to $\b \bbar$. Production at hadron colliders is therefore predominantly through $\g \g$ fusion, as implemented in process 149. The two main free parameters are the techni-$\eta$ mass and the decay constant $F_{\pi}$. The latter appears inversely quadratically in all the partial widths. Also the total cross section is affected, since the cross section is proportional to the $\g \g$ partial width. $F_{\pi}$ is stored in {\tt PARP(46)} and has the default value 123 GeV, which is the number predicted in some models. \subsection{Main Processes by Machine} In the previous section we have already commented on which processes have limited validity, or have different meanings (according to conventional terminology) in different contexts. Let us just repeat a few of the main points to be remembered for different machines. \subsubsection{$\ee$ collisions} The main annihilation process is number 1, $\ee \to \Z^0$, where in fact the full $\gamma^*/\Z^0$ interference structure is included. This process can be used, with some confidence, for c.m. energies from about 4 GeV upwards, i.e. at DORIS/CESR, PETRA/PEP, TRISTAN, LEP, and any future linear colliders. (To get below 10 GeV, you have to change {\tt PARP(2)}, however.) This is the default process obtained when {\tt MSEL=1}, i.e. when you do not change anything yourself. Process 141 contains a $\Z'^0$, including full interference with the standard $\gammaZ$. With the value {\tt MSTP(44)=4} in fact one is back at the standard $\gammaZ$ structure, i.e. the $\Z'^0$ piece has been switched off. Even so, this process may be useful, since it can simulate e.g. $\ee \to \H^0 \A^0$. Since the $\H^0$ may in its turn decay to $\Z^0 \Z^0$, a decay channel of the ordinary $\Z^0$ to $\H^0 \A^0$, although physically correct, would be technically confusing. In particular, it would be messy to set the original $\Z^0$ to decay one way and the subsequent ones another. So, in this sense, the $\Z'^0$ could be used as a copy of the ordinary $\Z^0$, but with a distinguishable label. The process $\ee \to \Upsilon$ does not exist as a separate process in {\Py}, but can be simulated by using {\tt LUONIA}, see section \ref{ss:oniadecays}. At LEP 2 and even higher energy machines, the simple $s$-channel process 1 will lose out to other processes, such as $\ee \to \Z^0 \Z^0$ and $\ee \to \W^+ \W^-$, i.e. processes 22 and 25. The former process in fact includes the structure $\ee \to (\gammaZ)(\gammaZ)$, which means that the cross section is singular if either of the two $\gammaZ$ masses is allowed to vanish. A mass cut therefore needs to be introduced, and is actually also used in other processes, such as $\ee \to \W^+ \W^-$. For practical applications, both with respect to cross sections and to event shapes, it is imperative to include initial-state radiation effects. Therefore {\tt MSTP(11)=1} is the default, wherein exponentiated electron-inside-electron distributions are used to give the momentum of the actually interacting electron. By radiative corrections to process 1, such processes as $\ee \to \gamma \Z^0$ are therefore automatically generated. If process 19 were to be used at the same time, this would mean that radiation were to be double-counted. In the alternative {\tt MSTP(11)=0}, electrons are assumed to deposit their full energy in the hard process, i.e. initial-state QED radiation is not included. This option is very useful, since it often corresponds to the `ideal' events that one wants to correct back to. Resolved electrons also means that one may have interactions between photons. This opens up the whole field of $\gamma\gamma$ processes, which is described in section \ref{sss:photoprodclass}. In particular, with {\tt MSTP(12)=1} photons may be resolved, i.e. photons need not only interact point-like, but can also interact like a hadron with a partonic substructure. The whole menagerie of hadron--hadron collider processes can then be accessed, with the exception of `minimum bias' type processes, including elastic and diffractive scattering. One day, there may exist models also for this. The thrust of the {\PyJe} programs is towards processes that involve hadron production, one way or another. Because of generalizations from other areas, also a few completely non-hadronic processes are available. These include Bhabha scattering, $\ee \to \ee$ in process 10, and photon pair production, $\ee \to \gamma \gamma$ in process 18. However, note that the precision that could be expected in a {\Py} simulation of those processes is certainly far less than that of dedicated programs. For one thing, electroweak loop effects are not included. For another, nowhere is the electron mass taken into account, which means that explicit cut-offs at some minimum $\pT$ are always necessary. \subsubsection{Lepton--hadron collisions} The issue of applications to $\e\p$ colliders has been covered in a recent report \cite{Sjo92b}. The default process for a lepton--hadron collider is deep inelastic scattering, $\ell \q \to \ell' \q'$, of process 10. This includes $\gamma^0/\Z^0/\W^{\pm}$ exchange, with full interference, as described in section \ref{sss:DISclass}. Radiation off the incoming lepton leg is included by {\tt MSTP(11)=1} and off the outgoing one by {\tt MSTJ(41)=2} (both are default). Note that both QED and QCD radiation (off the $\e$ and the $\q$ legs, respectively) are allowed to modify the $x$ and $Q^2$ values of the process, while the conventional approach in the literature is to allow only the former. Therefore an option (on by default) has been added to preserve these values by a post-facto rescaling, {\tt MSTP(23)=1}. In terms of cross sections, a more important set of processes are related to photoproduction, either with a point-like or with a resolved photon, see section \ref{sss:photoprodclass}. A complete description of photoproduction is available \cite{Sch93,Sch93a}, but needs three separate runs for the three distinct behaviours of a photon: point-like, VMD resolved and anomalous resolved. \subsubsection{Hadron--hadron collisions} The default is to include QCD jet production by $2 \to 2$ processes, see section \ref{sss:QCDjetclass}. Since the differential cross section is divergent for $\pT \to 0$, a lower cut-off has to be introduced. Normally that cut-off is given by the user-set $\pTmin$ value in {\tt CKIN(3)}. If {\tt CKIN(3)} is chosen smaller than a given value of the order of 2 GeV (see {\tt PARP(81)} and {\tt PARP(82)}), then low-$\pT$ events are also switched on. The jet cross section is regularized at low $\pT$, so as to obtain a smooth joining between the high-$\pT$ and the low-$\pT$ descriptions, see further section \ref{ss:multint}. As {\tt CKIN(3)} is varied, the jump from one scenario to another is abrupt, in terms of cross section: in a high-energy hadron collider, the cross section for jets down to a $\pTmin$ scale of a few GeV can well reach values much larger than the total inelastic, non-diffractive cross section. Clearly this is nonsense; therefore either $\pTmin$ should be picked so large that the jet cross section be only a fraction of the total one, or else one should select $\pTmin = 0$ and make use of the full description. If one switches to {\tt MSEL=2}, also elastic and diffractive processes are switched on, see section \ref{sss:minbiasclass}. However, the simulation of these processes is fairly primitive, and should not be used for dedicated studies, but only to estimate how much they may contaminate the class of non-diffractive minimum bias events. Most processes can be simulated in hadron colliders, since the bulk of {\Py} processes can be initiated by quarks or gluons. However, there are limits. Currently we include no photon or lepton parton distributions, which means that a process like $\gamma \q \to \gamma \q$ is not accessible. Further, the possibility of having $\Z^0$ and $\W^{\pm}$ interacting in processes such as 71--77 has been hardwired process by process, and does not mean that there is a generic treatment of $\Z^0$ and $\W^{\pm}$ distributions. The emphasis in the hadron--hadron process description is on high energy hadron colliders. The program can be used also at fixed-target energies, but the multiple interaction model for underlying events then breaks down and should not be used. The limit of applicability is somewhere at around 100 GeV. Below that, one is also recommended to change to {\tt MSTP(92)=3}, to obtain a reasonable amount of beam remnant particle production in the absence of multiple interactions. \clearpage \section{The PYTHIA Program Elements} In the previous two sections, the physics processes and the event-generation schemes of {\Py} have been presented. Here, finally, the event-generation routines and the common block variables are described. However, routines and variables related to initial- and final-state showers, beam remnants and underlying events, and fragmentation and decay are relegated to subsequent sections on these topics. Further, for historical reasons, many adjustable coupling constants are found in the {\tt LUDAT1} common block in {\Je}, rather than somewhere in the {\Py} common blocks; these parameters are described in section \ref{ss:coupcons}. In the presentation in this section, information less important for an efficient use of {\Py} has been put closer to the end. We therefore begin with the main event generation routines, and follow this by the main common block variables. It is useful to distinguish three phases in a normal run with {\Py}. In the first phase, the initialization, the general character of the run is determined. At a minimum, this requires the specification of the incoming hadrons and the energies involved. At the discretion of the user, it is also possible to select specific final states, and to make a number of decisions about details in the subsequent generation. This step is finished by a {\tt PYINIT} call, at which time several variables are initialized in accordance with the values set. The second phase consists of the main loop over the number of events, with each new event being generated by a {\tt PYEVNT} call. This event may then be analysed, using information stored in some common blocks, and the statistics accumulated. In the final phase, results are presented. This may often be done without the invocation of any {\Py} routines. From {\tt PYSTAT}, however, it is possible to obtain a useful list of cross sections for the different subprocesses. \subsection{The Main Subroutines} \label{ss:PYTmainroutines} There are two routines that you must know: {\tt PYINIT} for initialization and {\tt PYEVNT} for the subsequent generation of each new event. In addition, the cross section and other kinds of information available with {\tt PYSTAT} are frequently useful. The other three routines described here, {\tt PYFRAM}, {\tt PYKCUT}, and {\tt PYEVWT}, are of more specialized interest. \drawbox{CALL PYINIT(FRAME,BEAM,TARGET,WIN)}\label{p:PYINIT} \begin{entry} \itemc{Purpose:} to initialize the generation procedure. \iteme{FRAME :} a character variable used to specify the frame of the experiment. Upper-case and lower-case letters may be freely mixed. \begin{subentry} \iteme{= 'CMS' :} colliding beam experiment in c.m. frame, with beam momentum in $+z$ direction and target momentum in $-z$ direction. \iteme{= 'FIXT' :} fixed-target experiment, with beam particle momentum pointing in $+z$ direction. \iteme{= 'USER' :} full freedom to specify frame by giving beam momentum in {\tt P(1,1)}, {\tt P(1,2)} and {\tt P(1,3)} and target momentum in {\tt P(2,1)}, {\tt P(2,2)} and {\tt P(2,3)} in common block {\tt LUJETS}. Particles are assumed on the mass shell, and energies are calculated accordingly. \iteme{= 'FOUR' :} as {\tt 'USER'}, except also energies should be specified, in {\tt P(1,4)} and {\tt P(2,4)}, respectively. The particles need not be on the mass shell; effective masses are calculated from energy and momentum. (But note that numerical precision may suffer; if you know the masses the option {\tt 'FIVE'} below is preferrable.) \iteme{= 'FIVE' :} as {\tt 'USER'}, except also energies and masses should be specified, i.e the full momentum information in {\tt P(1,1) - P(1,5)} and {\tt P(2,1) - P(2,5)} should be given for beam and target, respectively. Particles need not be on the mass shell. Space-like virtualities should be stored as $-\sqrt{-m^2}$. Four-momentum and mass information must match. \iteme{= 'NONE' :} there will be no initialization of any processes, but only of resonance widths and a few other process-independent variables. Subsequent to such a call, {\tt PYEVNT} cannot be used to generate events, so this option is mainly intended for those who will want to construct their own events afterwards, but still want to have access to some of the {\Py} facilities. In this option, the {\tt BEAM}, {\tt TARGET} and {\tt WIN} arguments are dummy. \end{subentry} \iteme{BEAM, TARGET :} character variables to specify beam and target particles. Upper-case and lower-case letters may be freely mixed. An antiparticle may be denoted either by `$\sim$' or `bar' at the end of the name. It is also possible to leave out the underscore (`\_') directly after `nu' in neutrino names, and the charge for proton and neutron. \begin{subentry} \iteme{= 'e-' :} electron. \iteme{= 'e+' :} positron. \iteme{= 'nu\_e' :} $\nu_{\e}$. \iteme{= 'nu\_e$\sim$' :} $\br{\nu}_{\e}$. \iteme{= 'mu-' :} $\mu^-$. \iteme{= 'mu+' :} $\mu^+$. \iteme{= 'nu\_mu' :} $\nu_{\mu}$. \iteme{= 'nu\_mu$\sim$' :} $\br{\nu}_{\mu}$. \iteme{= 'tau-' :} $\tau^-$. \iteme{= 'tau+' :} $\tau^+$. \iteme{= 'nu\_tau' :} $\nu_{\tau}$. \iteme{= 'nu\_tau$\sim$' :} $\br{\nu}_{\tau}$. \iteme{= 'gamma' :} photon (real, i.e. on the mass shell). \iteme{= 'pi0' :} $\pi^0$. \iteme{= 'pi+' :} $\pi^+$. \iteme{= 'pi-' :} $\pi^-$. \iteme{= 'n0' :} neutron. \iteme{= 'n$\sim$0' :} antineutron. \iteme{= 'p+' :} proton. \iteme{= 'p$\sim$-' :} antiproton. \iteme{= 'Lambda0' :} $\Lambda$ baryon. \iteme{= 'Sigma-' :} $\Sigma^-$ baryon. \iteme{= 'Sigma0' :} $\Sigma^0$ baryon. \iteme{= 'Sigma+' :} $\Sigma^+$ baryon. \iteme{= 'Xi-' :} $\Xi^-$ baryon. \iteme{= 'Xi0' :} $\Xi^0$ baryon. \iteme{= 'Omega-' :} $\Omega^-$ baryon. \iteme{= 'pomeron' :} the pomeron $\pomeron$; since pomeron parton distribution functions have not been defined this option can not be used currently. \iteme{= 'reggeon' :} the reggeon $\reggeon$, with comments as for the pomeron above. \end{subentry} \iteme{WIN :} related to energy of system, exact meaning depends on {\tt FRAME}. \begin{subentry} \iteme{FRAME='CMS' :} total energy of system (in GeV). \iteme{FRAME='FIXT' :} momentum of beam particle (in GeV/$c$). \iteme{FRAME='USER' :} dummy (information is taken from the {\tt P} vectors, see above). \end{subentry} \end{entry} \drawbox{CALL PYEVNT}\label{p:PYEVNT} \begin{entry} \itemc{Purpose:} to generate one event of the type specified by the {\tt PYINIT} call. (This is the main routine, which calls a number of other routines for specific tasks.) \end{entry} \drawbox{CALL PYSTAT(MSTAT)}\label{p:PYSTAT} \begin{entry} \itemc{Purpose:} to print out cross-sections statistics, decay widths, branching ratios, status codes and parameter values. {\tt PYSTAT} may be called at any time, after the {\tt PYINIT} call, e.g. at the end of the run, or not at all. \iteme{MSTAT :} specification of desired information. \begin{subentry} \iteme{= 1 :} prints a table of how many events of the different kinds that have been generated and the corresponding cross sections. All numbers already include the effects of cuts required by the user in {\tt PYKCUT}. \iteme{= 2 :} prints a table of the resonances defined in the program, with their particle codes (KF), and all allowed decay channels. (If the number of generations in {\tt MSTP(1)} is 3, however, channels involving fourth-generation particles are not displayed.) For each decay channel is shown the sequential channel number (IDC) of the {\Je} decay tables, the partial decay width, branching ratio and effective branching ratio (in the event some channels have been excluded by the user). \iteme{= 3 :} prints a table with the allowed hard interaction flavours {\tt KFIN(I,J)} for beam and target particles. \iteme{= 4 :} prints a table of the kinematical cuts {\tt CKIN(I)} set by the user in the current run. \iteme{= 5 :} prints a table with all the values of the status codes {\tt MSTP(I)} and the parameters {\tt PARP(I)} used in the current run. \end{subentry} \end{entry} \drawbox{CALL PYFRAM(IFRAME)}\label{p:PYFRAM} \begin{entry} \itemc{Purpose:} to transform an event listing between different reference frames, if so desired. \iteme{IFRAME :} specification of frame the event is to be boosted to. \begin{subentry} \iteme{= 1 :} frame specified by user in the {\tt PYINIT} call. \iteme{= 2 :} c.m. frame of incoming particles. \iteme{= 3 :} hadronic c.m. frame of lepton--hadron interaction events. Mainly intended for deep inelastic scattering, but can also be used in photoproduction. Note that both the lepton and any photons radiated off the lepton remain in the event listing, and have to be removed separately if you only want to study the hadronic subsystem. \end{subentry} \end{entry} \drawbox{CALL PYKCUT(MCUT)}\label{p:PYKCUT} \begin{entry} \itemc{Purpose:} to enable you to reject a given set of kinematic variables at an early stage of the generation procedure (before evaluation of cross sections), so as not to spend unnecessary time on the generation of events that are not wanted. The routine will not be called unless you require is by setting {\tt MSTP(141)=1}, and never if `minimum-bias'-type events (including elastic and diffractive scattering) are to be generated as well. A dummy routine {\tt PYKCUT} is included in the program file, so as to avoid unresolved external references when the routine is not used. \iteme{MCUT :} flag to signal effect of user-defined cuts. \begin{subentry} \iteme{= 0 :} event is to be retained and generated in full. \iteme{= 1 :} event is to be rejected and a new one generated. \end{subentry} \itemc{Remark :} at the time of selection, several variables in the {\tt MINT} and {\tt VINT} arrays in the {\tt PYINT1} common block contain information that can be used to make the decision. The routine provided in the program file explicitly reads the variables that have been defined at the time {\tt PYKCUT} is called, and also calculates some derived quantities. The information available includes subprocess type ISUB, $E_{\mrm{cm}}$, $\hat{s}$, $\hat{t}$, $\hat{u}$, $\hat{p}_{\perp}$, $x_1$, $x_2$, $x_F$, $\tau$, $y$, $\tau'$, $\cos\hat{\theta}$, and a few more. Some of these may not be relevant for the process under study, and are then set to zero. \end{entry} \drawbox{CALL PYEVWT(WTXS)}\label{p:PYEVWT} \begin{entry} \itemc{Purpose:} to allow you to reweight event cross sections, by process type and kinematics of the hard scattering. There exists two separate modes of usage, described in the following. \\ For {\tt MSTP(142)=1}, it is assumed that the cross section of the process is correctly given by default in {\Py}, but that one wishes to generate events biased to a specific region of phase space. While the {\tt WTXS} factor therefore multiplies the na\"{\i}ve cross section in the choice of subprocess type and kinematics, the produced event comes with a compensating weight {\tt PARI(10)=1./WTXS}, which should be used when filling histograms etc. In the {\tt PYSTAT(1)} table, the cross sections are unchanged (up to statistical errors) compared with the standard cross sections, but the relative composition of events may be changed and need no longer be in proportion to relative cross sections. A typical example of this usage is if one wishes to enhance the production of high-$\pT$ events; then a weight like {\tt WTXS}$=(\pT/p_{\perp 0})^2$ (with $p_{\perp 0}$ some fixed number) might be appropriate. \\ For {\tt MSTP(142)=2}, on the other hand, it is assumed that the true cross section is really to be modifed by the multiplicative factor {\tt WTXS}. The generated events therefore come with unit weight, just as usual. This option is really equivalent to replacing the basic cross sections coded in {\Py}, but allows more flexibility: no need to recompile the whole of {\Py}. \\ The routine will not be called unless {\tt MSTP(142)}$\geq 1$, and never if `minimum-bias'-type events (including elastic and diffractive scattering) are to be generated as well. Further, cross sections for additional multiple interactions or pile-up events are never affected. A dummy routine {\tt PYEVWT} is included in the program file, so as to avoid unresolved external references when the routine is not used. \iteme{WTXS:} multiplication factor to ordinary event cross section; to be set (by you) in {\tt PYEVWT} call. \itemc{Remark :} at the time of selection, several variables in the {\tt MINT} and {\tt VINT} arrays in the {\tt PYINT1} common block contain information that can be used to make the decision. The routine provided in the program file explicitly reads the variables that have been defined at the time {\tt PYEVWT} is called, and also calculates some derived quantities. The given list of information includes subprocess type ISUB, $E_{\mrm{cm}}$, $\hat{s}$, $\hat{t}$, $\hat{u}$, $\hat{p}_{\perp}$, $x_1$, $x_2$, $x_F$, $\tau$, $y$, $\tau'$, $\cos\hat{\theta}$, and a few more. Some of these may not be relevant for the process under study, and are then set to zero. \itemc{Warning:} the weights only apply to the hard scattering subprocesses. There is no way to reweight the shape of initial- and final-state showers, fragmentation, or other aspects of the event. \end{entry} \subsection{Switches for Event Type and Kinematics Selection} \label{ss:PYswitchkin} By default, if {\Py} is run for a hadron collider, only QCD $2 \to 2$ processes are generated, composed of hard interactions above $\pTmin=${\tt PARP(81)}, with low-$\pT$ processes added on so as to give the full (parametrized) inelastic, non-diffractive cross section. In an $\ee$ collider, $\gammaZ$ production is the default, and in an $\ep$ one it is deep inelastic scattering. With the help of the common block {\tt PYSUBS}, it is possible to select the generation of another process, or combination of processes. It is also allowed to restrict the generation to specific incoming partons/particles at the hard interaction. This often automatically also restricts final-state flavours but, in processes such as resonance production or QCD/QED production of new flavours, switches in the {\Je} program may be used to this end; see section \ref{ss:parapartdat}. The {\tt CKIN} array may be used to impose specific kinematics cuts. You should here be warned that, if kinematical variables are too strongly restricted, the generation time per event may become very long. In extreme cases, where the cuts effectively close the full phase space, the event generation may run into an infinite loop. The generation of $2 \to 1$ resonance production is performed in terms of the $\hat{s}$ and $y$ variables, and so the ranges {\tt CKIN(1) - CKIN(2)} and {\tt CKIN(7) - CKIN(8)} may be arbitrarily restricted without a significant loss of speed. For $2 \to 2$ processes, $\cos\hat{\theta}$ is added as a third generation variable, and so additionally the range {\tt CKIN(27) - CKIN(28)} may be restricted without any danger. Effects from initial- and final-state radiation are not included, since they are not known at the time the kinematics at the hard interaction is selected. The sharp kinematical cut-offs that can be imposed on the generation process are therefore smeared, both by QCD radiation and by fragmentation. A few examples of such effects follow. \begin{Itemize} \item Initial-state radiation implies that each of the two incoming partons has a non-vanishing $\pT$ when they interact. The hard scattering subsystem thus receives a net transverse boost, and is rotated with respect to the beam directions. In a $2 \to 2$ process, what typically happens is that one of the scattered partons receives an increased $\pT$, while the $\pT$ of the other parton is reduced. \item Since the initial-state radiation machinery assigns space-like virtualities to the incoming partons, the definitions of $x$ in terms of energy fractions and in terms of momentum fractions no longer coincide, and so the interacting subsystem may receive a net longitudinal boost compared with na\"{\i}ve expectations, as part of the parton-shower machinery. \item Initial-state radiation gives rise to additional jets, which in extreme cases may be mistaken for either of the jets of the hard interaction. \item Final-state radiation gives rise to additional jets, which smears the meaning of the basic $2 \to 2$ scattering. The assignment of soft jets is not unique. The energy of a jet becomes dependent on the way it is identified, e.g. what jet cone size is used. \item The beam remnant description assigns primordial $k_{\perp}$ values, which also gives a net $\pT$ shift of the hard-interaction subsystem; except at low energies this effect is overshadowed by initial-state radiation, however. Beam remnants may also add further activity under the `perturbative' event. \item Fragmentation will further broaden jet profiles, and make jet assignments and energy determinations even more uncertain. \end{Itemize} In a study of events within a given window of experimentally defined variables, it is up to you to leave such liberal margins that no events are missed. In other words, cuts have to be chosen such that a negligible fraction of events migrate from outside the simulated region to inside the interesting region. Often this may lead to low efficiency in terms of what fraction of the generated events are actually of interest to you. See also section \ref{ss:PYTstarted}. In addition to the variables found in {\tt PYSUBS}, also those in the {\tt PYPARS} common block may be used to select exactly what one wants to have simulated. These possibilities will be described in the following subsection. The notation used above and in the following is that `$\hat{~}$' denotes internal variables in the hard scattering subsystem, while `$^*$' is for variables in the c.m. frame of the event as a whole. \drawbox{COMMON/PYSUBS/MSEL,MSUB(200),KFIN(2,-40:40),CKIN(200)}% \label{p:PYSUBS} \begin{entry} \itemc{Purpose:} to allow the user to run the program with any desired subset of processes, or restrict flavours or kinematics. If the default values, denoted below by (D=\ldots), are not satisfactory, they must be changed before the {\tt PYINIT} call. \boxsep \iteme{MSEL :}\label{p:MSEL} (D=1) a switch to select between full user control and some preprogrammed alternatives. \begin{subentry} \iteme{= 0 :} desired subprocesses have to be switched on in {\tt MSUB}, i.e. full user control. \iteme{= 1 :} depending on incoming particles, different alternatives are used. \\ Lepton--lepton: $\Z$ or $\W$ production (ISUB = 1 or 2). \\ Lepton--hadron: deep inelastic scattering (ISUB = 10). \\ Hadron--hadron: QCD high-$\pT$ processes (ISUB = 11, 12, 13, 28, 53, 68); additionally low-$\pT$ production if {\tt CKIN(3)}$<${\tt PARP(81)} or {\tt PARP(82)}, depending on {\tt MSTP(82)} (ISUB = 95). If low-$\pT$ is switched on, the other {\tt CKIN} cuts are not used. \\ A resolved photon ({\tt MSTP(14)}$\geq 1$) counts as hadron, except that an anomalous photon ({\tt MSTP(14)=3}) can not have low-$\pT$ interactions. When the photon is not resolved, the following cases are possible. \\ Photon--lepton: Compton scattering (ISUB = 34). \\ Photon--hadron: photon-parton scattering (ISUB = 33, 34, 54). \\ Photon--photon: fermion pair production (ISUB = 58). \iteme{= 2 :} as {\tt MSEL = 1} for lepton--lepton, lepton--hadron and unresolved photons. For hadron--hadron (including resolved photons) all QCD processes, including low-$\pT$, single and double diffractive and elastic scattering, are included (ISUB = 11, 12, 13, 28, 53, 68, 91, 92, 93, 94, 95). The {\tt CKIN} cuts are here not used. \iteme{= 4 :} charm ($\c\cbar$) production with massive matrix elements (ISUB = 81, 82, 84, 85). \iteme{= 5 :} bottom ($\b\bbar$) production with massive matrix elements (ISUB = 81, 82, 84, 85). \iteme{= 6 :} top ($\t\tbar$) production with massive matrix elements (ISUB = 81, 82, 84, 85). \iteme{= 7 :} low ($\lrm\br{\lrm}$) production with massive matrix elements (ISUB = 81, 82, 84, 85). \iteme{= 8 :} high ($\hrm\br{\hrm}$) production with massive matrix elements (ISUB = 81, 82, 84, 85). \iteme{= 10 :} prompt photons (ISUB = 14, 18, 29). \iteme{= 11 :} $\Z^0$ production (ISUB = 1). \iteme{= 12 :} $\W^{\pm}$ production (ISUB = 2). \iteme{= 13 :} $\Z^0$ + jet production (ISUB = 15, 30). \iteme{= 14 :} $\W^{\pm}$ + jet production (ISUB = 16, 31). \iteme{= 15 :} pair production of different combinations of $\gamma$, $\Z^0$ and $\W^{\pm}$ (except $\gamma\gamma$; see {\tt MSEL} = 10) (ISUB = 19, 20, 22, 23, 25). \iteme{= 16 :} $\H^0$ production (ISUB = 3, 102, 103, 123, 124). \iteme{= 17 :} $\H^0 \Z^0$ or $\H^0 \W^{\pm}$ (ISUB = 24, 26). \iteme{= 18 :} $\H^0$ production, combination relevant for $\ee$ annihilation (ISUB = 24, 103, 123, 124). \iteme{= 19 :} $\H^0$, $\H'^0$ and $\A^0$ production, excepting pair production (ISUB = 24, 103, 123, 124, 153, 158, 171, 173, 174, 176, 178, 179). \iteme{= 21 :} $\Z'^0$ production (ISUB = 141). \iteme{= 22 :} $\W'^{\pm}$ production (ISUB = 142). \iteme{= 23 :} $\H^{\pm}$ production (ISUB = 143). \iteme{= 24 :} $\R^0$ production (ISUB = 144). \iteme{= 25 :} $\L_{\Q}$ (leptoquark) production (ISUB = 145, 162, 163, 164). \iteme{= 35:} single bottom production by $\W$ exchange (ISUB = 83). \iteme{= 36:} single top production by $\W$ exchange (ISUB = 83). \iteme{= 37:} single low production by $\W$ exchange (ISUB = 83). \iteme{= 38:} single high production by $\W$ exchange (ISUB = 83). \end{subentry} \boxsep \iteme{MSUB :}\label{p:MSUB} (D=200*0) array to be set when {\tt MSEL=0} (for {\tt MSEL}$\geq 1$ relevant entries are set in {\tt PYINIT}) to choose which subset of subprocesses to include in the generation. The ordering follows the ISUB code given in section \ref{ss:ISUBcode} (with comments as given there). \begin{subentry} \iteme{MSUB(ISUB) = 0 :} the subprocess is excluded. \iteme{MSUB(ISUB) = 1 :} the subprocess is included. \itemc{Note:} when {\tt MSEL=0}, the {\tt MSUB} values set by the user are never changed by {\Py}. If you want to combine several different `subruns', each with its own {\tt PYINIT} call, into one single run, it is up to you to remember not only to switch on the new processes before each new {\tt PYINIT} call, but also to switch off the old ones that are no longer desired. \end{subentry} \boxsep \iteme{KFIN(I,J) :}\label{p:KFIN} provides an option to selectively switch on and off contributions to the cross sections from the different incoming partons/particles at the hard interaction. In combination with the {\Je} resonance decay switches, this also allows you to set restrictions on flavours appearing in the final state. \begin{subentry} \iteme{I :} is 1 for beam side of event and 2 for target side. \iteme{J :} enumerates flavours according to the KF code; see section \ref{ss:codes}. \iteme{KFIN(I,J) = 0 :} the parton/particle is forbidden. \iteme{KFIN(I,J) = 1 :} the parton/particle is allowed. \itemc{Note:} By default, the following are switched on: $\d$, $\u$, $\s$, $\c$, $\b$, $\e^-$, $\nu_{\e}$, $\mu^-$, $\nu_{\mu}$, $\tau^-$, $\nu_{\tau}$, $\g$, $\gamma$, $\Z^0$, $\W^+$ and their antiparticles. In particular, top is off, and has to be switched on explicitly if needed. \end{subentry} \boxsep \iteme{CKIN :}\label{p:CKIN} kinematics cuts that can be set by you before the {\tt PYINIT} call, and that affect the region of phase space within which events are generated. Some cuts are `hardwired' while most are `softwired'. The hardwired ones are directly related to the kinematical variables used in the event selection procedure, and therefore have negligible effects on program efficiency. The most important of these are {\tt CKIN(1) - CKIN(8)}, {\tt CKIN(27) - CKIN(28)}, and {\tt CKIN(31) - CKIN(32)}. The softwired ones are most of the remaining ones, that cannot be fully taken into account in the kinematical variable selection, so that generation in constrained regions of phase space may be slow. In extreme cases the phase space may be so small that the maximization procedure fails to find any allowed points at all (although some small region might still exist somewhere), and therefore switches off some subprocesses, or aborts altogether. \iteme{CKIN(1), CKIN(2) :} (D=2.,-1.) range of allowed $\hat{m} = \sqrt{\hat{s}}$ values. If {\tt CKIN(2)}$ < 0.$, the upper limit is inactive. \iteme{CKIN(3), CKIN(4) :} (D=0.,-1.) range of allowed $\hat{p}_{\perp}$ values for hard $2 \to 2$ processes, with transverse momentum $\hat{p}_{\perp}$ defined in the rest frame of the hard interaction. If {\tt CKIN(4)}$ < 0.$, the upper limit is inactive. For processes that are singular in the limit $\hat{p}_{\perp} \to 0$ (see {\tt CKIN(6)}), {\tt CKIN(5)} provides an additional constraint. The {\tt CKIN(3)} and {\tt CKIN(4)} limits can also be used in $2 \to 1 \to 2$ processes. Here, however, the product masses are not known and hence are assumed to be vanishing in the event selection. The actual $\pT$ range for massive products is thus shifted downwards with respect to the nominal one. \iteme{CKIN(5) :} (D=1.) lower cut-off on $\hat{p}_{\perp}$ values, in addition to the {\tt CKIN(3)} cut above, for processes that are singular in the limit $\hat{p}_{\perp} \to 0$ (see {\tt CKIN(6)}). \iteme{CKIN(6) :} (D=1.) hard $2 \to 2$ processes, which do not proceed only via an intermediate resonance (i.e. are $2 \to 1 \to 2$ processes), are classified as singular in the limit $\hat{p}_{\perp} \to 0$ if either or both of the two final-state products has a mass $m < ${\tt CKIN(6)}. \iteme{CKIN(7), CKIN(8) :} (D=-10.,10.) range of allowed scattering subsystem rapidities $y = y^*$ in the c.m. frame of the event, where $y = (1/2) \ln(x_1/x_2)$. (Following the notation of this section, the variable should be given as $y^*$, but because of its frequent use, it was called $y$ in section \ref{ss:kinemtwo}.) \iteme{CKIN(9), CKIN(10) :} (D=-10.,10.) range of allowed (true) rapidities for the product with largest rapidity in a $2 \to 2$ or a $2 \to 1 \to 2$ process, defined in the c.m. frame of the event, i.e. $\max(y^*_3, y^*_4)$. Note that rapidities are counted with sign, i.e. if $y^*_3 = 1$ and $y^*_4 = -2$ then $\max(y^*_3, y^*_4) = 1$. \iteme{CKIN(11), CKIN(12) :} (D=-10.,10.) range of allowed (true) rapidities for the product with smallest rapidity in a $2 \to 2$ or a $2 \to 1 \to 2$ process, defined in the c.m. frame of the event, i.e. $\min(y^*_3, y^*_4)$. Consistency thus requires {\tt CKIN(11)}$\leq${\tt CKIN(9)} and {\tt CKIN(12)}$\leq${\tt CKIN(10)}. \iteme{CKIN(13), CKIN(14) :} (D=-10.,10.) range of allowed pseudorapidities for the product with largest pseudorapidity in a $2 \to 2$ or a $2 \to 1 \to 2$ process, defined in the c.m. frame of the event, i.e. $\max(\eta^*_3, \eta^*_4)$. Note that pseudorapidities are counted with sign, i.e. if $\eta^*_3 = 1$ and $\eta^*_4 = -2$ then $\max(\eta^*_3, \eta^*_4) = 1$. \iteme{CKIN(15), CKIN(16) :} (D=-10.,10.) range of allowed pseudorapidities for the product with smallest pseudorapidity in a $2 \to 2$ or a $2 \to 1 \to 2$ process, defined in the c.m. frame of the event, i.e. $\min(\eta^*_3, \eta^*_4)$. Consistency thus requires {\tt CKIN(15)}$\leq${\tt CKIN(13)} and {\tt CKIN(16)}$\leq${\tt CKIN(14)}. \iteme{CKIN(17), CKIN(18) :} (D=-1.,1.) range of allowed $\cos\theta^*$ values for the product with largest $\cos\theta^*$ value in a $2 \to 2$ or a $2 \to 1 \to 2$ process, defined in the c.m. frame of the event, i.e. $\max(\cos\theta^*_3,\cos\theta^*_4)$. \iteme{CKIN(19), CKIN(20) :} (D=-1.,1.) range of allowed $\cos\theta^*$ values for the product with smallest $\cos\theta^*$ value in a $2 \to 2$ or a $2 \to 1 \to 2$ process, defined in the c.m. frame of the event, i.e. $\min(\cos\theta^*_3,\cos\theta^*_4)$. Consistency thus requires {\tt CKIN(19)}$\leq${\tt CKIN(17)} and {\tt CKIN(20)}$\leq${\tt CKIN(18)}. \iteme{CKIN(21), CKIN(22) :} (D=0.,1.) range of allowed $x_1$ values for the parton on side 1 that enters the hard interaction. \iteme{CKIN(23), CKIN(24) :} (D=0.,1.) range of allowed $x_2$ values for the parton on side 2 that enters the hard interaction. \iteme{CKIN(25), CKIN(26) :} (D=-1.,1.) range of allowed Feynman-$x$ values, where $x_F = x_1 - x_2$. \iteme{CKIN(27), CKIN(28) :} (D=-1.,1.) range of allowed $\cos\hat{\theta}$ values in a hard $2 \to 2$ scattering, where $\hat{\theta}$ is the scattering angle in the rest frame of the hard interaction. \iteme{CKIN(31), CKIN(32) :} (D=2.,-1.) range of allowed $\hat{m}' = \sqrt{\hat{s}'}$ values, where $\hat{m}'$ is the mass of the complete three- or four-body final state in $2 \to 3$ or $2 \to 4$ processes (while $\hat{m}$, constrained in {\tt CKIN(1)} and {\tt CKIN(2)}, here corresponds to the one- or two-body central system). If {\tt CKIN(32)}$ < 0.$, the upper limit is inactive. \iteme{CKIN(35), CKIN(36) :} (D=0.,-1.) range of allowed $|\hat{t}| = - \hat{t}$ values in $2 \to 2$ processes. Note that for deep inelastic scattering this is nothing but the $Q^2$ scale, in the limit that initial- and final-state radiation is neglected. If {\tt CKIN(36)}$ < 0.$, the upper limit is inactive. \iteme{CKIN(37), CKIN(38) :} (D=0.,-1.) range of allowed $|\hat{u}| = - \hat{u}$ values in $2 \to 2$ processes. If {\tt CKIN(38)}$ < 0.$, the upper limit is inactive. \iteme{CKIN(41) - CKIN(44) :} (D=12.,-1.,12.,-1.) range of allowed mass values of the two (or one) resonances produced in a `true' $2 \to 2$ process, i.e. one not (only) proceeding through a single $s$-channel resonance ($2 \to 1 \to 2$). (These are the ones listed as $2 \to 2$ in the tables in section \ref{ss:ISUBcode}.) Only particles with a width above {\tt PARP(41)} are considered as bona fide resonances and tested against the {\tt CKIN} limits; particles with a smaller width are put on the mass shell without applying any cuts. The exact interpretation of the {\tt CKIN} variables depends on the flavours of the two produced resonances. \\ For two resonances like $\Z^0 \W^+$ (produced from $\f \f' \to \Z^0\W^+$), which are not identical and which are not each other's antiparticles, one has \\ {\tt CKIN(41)}$ < m_1 < ${\tt CKIN(42)}, and \\ {\tt CKIN(43)}$ < m_2 < ${\tt CKIN(44)}, \\ where $m_1$ and $m_2$ are the actually generated masses of the two resonances, and 1 and 2 are defined by the order in which they are given in the production process specification. \\ For two resonances like $\Z^0 \Z^0$, which are identical, or $\W^+ \W^-$, which are each other's antiparticles, one instead has \\ {\tt CKIN(41)}$ < \min(m_1,m_2) < ${\tt CKIN(42)}, and \\ {\tt CKIN(43)}$ < \max(m_1,m_2) < ${\tt CKIN(44)}. \\ In addition, whatever limits are set on {\tt CKIN(1)} and, in particular, on {\tt CKIN(2)} obviously affect the masses actually selected. \begin{subentry} \itemc{Note 1:} If {\tt MSTP(42)=0}, so that no mass smearing is allowed, the {\tt CKIN} values have no effect (the same as for particles with too narrow a width). \itemc{Note 2:} If {\tt CKIN(42)}$ < ${\tt CKIN(41)} it means that the {\tt CKIN(42)} limit is inactive; correspondingly, if {\tt CKIN(44)}$ < ${\tt CKIN(43)} then {\tt CKIN(44)} is inactive. \itemc{Note 3:} If limits are active and the resonances are identical, it is up to you to ensure that {\tt CKIN(41)}$\leq${\tt CKIN(43)} and {\tt CKIN(42)}$\leq${\tt CKIN(44)}. \itemc{Note 4:} For identical resonances, it is not possible to preselect which of the resonances is the lighter one; if, for instance, one $\Z^0$ is to decay to leptons and the other to quarks, there is no mechanism to guarantee that the lepton pair has a mass smaller than the quark one. \itemc{Note 5:} The {\tt CKIN} values are applied to all relevant $2 \to 2$ processes equally, which may not be what one desires if several processes are generated simultaneously. Some caution is therefore urged in the use of the {\tt CKIN(41) - CKIN(44)} values. Also in other respects, users are recommended to take proper care: if a $\Z^0$ is only allowed to decay into $\b\bbar$, for example, setting its mass range to be 2--8 GeV is obviously not a good idea. \itemc{Note 6:} In principle, the machinery should work for any $2 \to 2$ process with resonances in the final state, but so far it has only been checked for processes 22--26, so also from this point some caution is urged. \end{subentry} \iteme{CKIN(45) - CKIN(48) :} (D=12.,-1.,12.,-1.) range of allowed mass values of the two (or one) secondary resonances produced in a $2 \to 1 \to 2$ process (like $\g \g \to \H^0 \to \Z^0 \Z^0$) or even a $2 \to 2 \to 4$ (or 3) process (like $\q \qbar \to \Z^0 \H^0 \to \Z^0 \W^+ \W^-$). Note that these {\tt CKIN} values only affect the secondary resonances; the primary ones are constrained by {\tt CKIN(1)}, {\tt CKIN(2)} and {\tt CKIN(41) - CKIN(44)} (indirectly, of course, the choice of primary resonance masses affects the allowed mass range for the secondary ones). What is considered to be a resonance is defined by {\tt PARP(41)}; particles with a width smaller than this are automatically put on the mass shell. The description closely parallels the one given for {\tt CKIN(41) - CKIN(44)}. Thus, for two resonances that are not identical or each other's antiparticles, one has \\ {\tt CKIN(45)}$ < m_1 < ${\tt CKIN(46)}, and \\ {\tt CKIN(47)}$ < m_2 < ${\tt CKIN(48)}, \\ where $m_1$ and $m_2$ are the actually generated masses of the two resonances, and 1 and 2 are defined by the order in which they given in the decay channel specification in the program (see e.g. output from {\tt PYSTAT(2)} or {\tt LULIST(12)}). For two resonances that are identical or each other's antiparticles, one instead has \\ {\tt CKIN(45)}$ < \min(m_1,m_2) < ${\tt CKIN(46)}, and \\ {\tt CKIN(47)}$ < \max(m_1,m_2) < ${\tt CKIN(48)}. \begin{subentry} \itemc{Notes 1 - 5:} as for {\tt CKIN(41) - CKIN(44)}, with trivial modifications. \itemc{Note 6:} Setting limits on secondary resonance masses is possible in any of the channels of the allowed types (see above). However, so far only $\H^0 \to \Z^0 \Z^0$ and $\H^0 \to \W^+ \W^-$ have been fully implemented, such that an arbitrary mass range below the na\"{\i}ve mass threshold may be picked. For other possible resonances, any restrictions made on the allowed mass range are not reflected in the cross section; and further it is not recommendable to pick mass windows that make a decay on the mass shell impossible. These limitations will be relaxed in future versions. \end{subentry} \iteme{CKIN(51) - CKIN(56) :} (D=0.,-1.,0.,-1.,0.,-1.) range of allowed transverse momenta in a true $2 \to 3$ process. Currently two different alternatives are around. For subprocess 131, the $\pT$ of the first product (the $\Z^0$) is set by {\tt CKIN(3)} and {\tt CKIN(4)}, while for the quark and antiquark $\pT$'s one has \\ {\tt CKIN(51)}$<\min(p_{\perp \q},p_{\perp \qbar})<${\tt CKIN(52)}, and \\ {\tt CKIN(53)}$<\max(p_{\perp \q},p_{\perp \qbar})<${\tt CKIN(54)}. \\ Negative {\tt CKIN(52)} and {\tt CKIN(54)} values means that the corresponding limits are inactive. For subprocesses 121--124, and their $\H'^0$ and $\A^0$ equivalents (173, 174, 178, 179, 181, 182, 186, 187), {\tt CKIN(51) - CKIN(54)} again corresponds to $\pT$ ranges for scattered partons, but in order of appearance, i.e. {\tt CKIN(51) - CKIN(52)} for the parton scattered off the beam and {\tt CKIN(53) - CKIN(54)} for the one scattered off the target. {\tt CKIN(55)} and {\tt CKIN(56)} here sets $\pT$ limits for the third product, the $\H^0$, i.e. the {\tt CKIN(3)} and {\tt CKIN(4)} values have no effect for this process. Since the $\pT$ of the Higgs is not one of the primary variables selected, any constraints here may mean reduced Monte Carlo efficiency, while for these processes {\tt CKIN(51) - CKIN(54)} are `hardwired' and therefore do not cost anything. \end{entry} \subsection{The General Switches and Parameters} \label{ss:PYswitchpar} The {\tt PYPARS} common block contains the status code and parameters that regulate the performance of the program. All of them are provided with sensible default values, so that a novice user can neglect them, and only gradually explore the full range of possibilities. Some of the switches and parameters in {\tt PYPARS} will be described later, in the shower and beam remnants sections. \drawbox{COMMON/PYPARS/MSTP(200),PARP(200),MSTI(200),PARI(200)}% \label{p:PYPARS1} \begin{entry} \itemc{Purpose:} to give access to status code and parameters that regulate the performance of the program. If the default values, denoted below by (D=\ldots), are not satisfactory, they must in general be changed before the {\tt PYINIT} call. Exceptions, i.e. variables that can be changed for each new event, are denoted by (C). \iteme{MSTP(1) :}\label{p:MSTP} (D=3) maximum number of generations. Automatically set $\leq 4$. \iteme{MSTP(2) :} (D=1) calculation of $\alphas$ at hard interaction, in the routine {\tt ULALPS}. \begin{subentry} \iteme{= 0 :} $\alphas$ is fixed at value {\tt PARU(111)}. \iteme{= 1 :} first-order running $\alphas$. \iteme{= 2 :} second-order running $\alphas$. \end{subentry} \iteme{MSTP(3) :} (D=2) selection of $\Lambda$ value in $\alphas$ for {\tt MSTP(2)}$\geq 1$. \begin{subentry} \iteme{= 1 :} $\Lambda$ is given by {\tt PARP(1)} for hard interactions, by {\tt PARP(61)} for space-like showers and by {\tt PARJ(81)} for time-like ones. This $\Lambda$ is assumed to be valid for 5 flavours; for the hard interaction the number of flavours assumed can be changed by {\tt MSTU(112)}. \iteme{= 2 :} $\Lambda$ value is chosen according to the parton-distribution-function para\-metri\-zations, i.e. $\Lambda$ = 0.20 GeV for EHLQ1, = 0.29 GeV for EHLQ2, = 0.20 GeV for DO1, = 0.40 GeV for DO2, = 0.213 GeV for CTEQ2M, = 0.208 GeV for CTEQ2MS, = 0.208 GeV for CTEQ2MF, = 0.322 GeV for CTEQ2ML, = 0.190 GeV for CTEQ2L, = 0.235 GeV for CTEQ2D, and similarly for parton-distribution functions in the {\sc Pdflib} library (cf. ({\tt MSTP(51)}, {\tt MSTP(52)}). The choice is always based on the proton parton-distribution set selected, i.e. is unaffected by pion and photon parton-distribution selection. All the $\Lambda$ values above are assumed to refer to 4 flavours, and {\tt MSTU(112)} is set accordingly. This $\Lambda$ value is used both for the hard scattering and the initial- and final-state radiation. The ambiguity in the choice of the $Q^2$ argument still remains (see {\tt MSTP(32)}, {\tt MSTP(64)} and {\tt MSTJ(44)}). This $\Lambda$ value is used also for {\tt MSTP(57)=0}, but the sensible choice here would be to use {\tt MSTP(2)=0} and have no initial- or final-state radiation. \end{subentry} \iteme{MSTP(4) :} (D=0) treatment of the Higgs sector, predominantly the neutral one. \begin{subentry} \iteme{= 0 :} the $\H^0$ is given the Standard Model Higgs couplings, while $\H'^0$ and $\A^0$ couplings should be set by the user in {\tt PARU(171) - PARU(175)} and {\tt PARU(181) - PARU(185)}, respectively. \iteme{= 1 :} the user should set couplings for all three Higgses, for the $\H^0$ in {\tt PARU(161) - PARU(165)}, and for the $\H'^0$ and $\A^0$ as above. \iteme{= 2 :} the mass of $\H^0$ in {\tt PMAS(25,1)} and the $\tan\beta$ value in {\tt PARU(141)} are used to derive $\H'^0$, $\A^0$ and $\H^{\pm}$ masses, and $\H^0$, $\H'^0$, $\A^0$ and $\H^{\pm}$ couplings, using the relations of the Minimal Supersymmetric extension of the Standard Model at Born level \cite{Gun90}. Existing masses and couplings are overwritten by the derived values. See section \ref{sss:extneutHclass} for discussion on parameter constraints. \iteme{= 3:} as {\tt =2}, but using relations at the one-loop level. This option is not yet implemented. \end{subentry} \iteme{MSTP(5) :} (D=0) presence of anomalous couplings in processes. \begin{subentry} \iteme{= 0 :} absent. \iteme{$\geq$1 :} present, wherever implemented. See section \ref{sss:ancoupclass} for further details. \end{subentry} \iteme{MSTP(6) :} (D=0) usage of the fourth-generation fermions to simulate other fermion kinds. \begin{subentry} \iteme{= 0 :} none, i.e. can be used as a standard fourth generation. \iteme{= 1 :} excited fermions, as present in compositeness scenarios; see section \ref{sss:qlstarclass}. \end{subentry} \iteme{MSTP(7) :} (D=0) choice of heavy flavour in subprocesses 81--85. Does not apply for {\tt MSEL=4-8}, where the MSEL value always takes precedence. \begin{subentry} \iteme{= 0 :} for processes 81--84 (85) the `heaviest' flavour allowed for gluon (photon) splitting into a quark--antiquark (fermion--antifermion) pair, as set in the {\tt MDME} array. Note that `heavy' is defined as the one with largest KF code, so that leptons take precedence if they are allowed. \iteme{= 1 - 8 :} pick this particular quark flavour; e.g., {\tt MSTP(7)=6} means that top will be produced. \iteme{= 11 - 18 :} pick this particular lepton flavour. Note that neutrinos are not possible, i.e. only 11, 13, 15 and 17 are meaningful alternatives. Lepton pair production can only occur in process 85, so if any of the other processes have been switched on they are generated with the same flavour as would be obtained in the option {\tt MSTP(7)=0}. \end{subentry} \iteme{MSTP(11) :} (D=1) use of electron parton distribution in $\ee$ and $\ep$ interactions. \begin{subentry} \iteme{= 0 :} no, i.e. electron carries the whole beam energy. \iteme{= 1 :} yes, i.e. electron carries only a fraction of beam energy in agreement with next-to-leading electron parton-distribution function, thereby including the effects of initial-state bremsstrahlung. \end{subentry} \iteme{MSTP(12) :} (D=0) use of $\e^-$ (`sea', i.e. from $\e \to \gamma \to \e$), $\e^+$, quark and gluon distribution functions inside an electron. \begin{subentry} \iteme{= 0 :} off. \iteme{= 1 :} on, provided that {\tt MSTP(11)}$\geq 1$. Quark and gluons distributions are obtained by numerical convolution of the photon content inside an electron (as given by the bremsstrahlung spectrum of {\tt MSTP(11)=1}) with the quark and gluon content inside a photon. The required numerical precision is set by {\tt PARP(14)}. Since the need for numerical integration makes this option somewhat more time-consuming than ordinary parton-distribution evaluation, one should only use it when studying processes where it is needed. \end{subentry} \iteme{MSTP(13) :} (D=1) choice of $Q^2$ range over which electrons are assumed to radiate photons; affects normalization of $\e^-$ (sea), $\e^+$, $\gamma$, quark and gluon distributions inside an electron. \begin{subentry} \iteme{= 1 :} range set by $Q^2$ argument of parton-distribution-function call, i.e. by $Q^2$ scale of the hard interaction. Therefore parton distributions are proportional to $\ln(Q^2/m_e^2)$. This is normally most appropriate for $\ee$ annihilation. \iteme{= 2 :} range set by the user-determined $Q_{\mmax}^2$, given in {\tt PARP(13)}. Parton distributions are assumed to be proportional to $\ln((Q_{\mmax}^2/m_e^2)(1-x)/x^2)$. This is normally most appropriate for photoproduction, where the electron is supposed to go undetected, i.e. scatter less than $Q_{\mmax}^2$. \itemc{Note:} the choice of effective range is especially touchy for the quark and gluon distributions. An (almost) on-the-mass-shell photon has a VMD piece that dies away for a virtual photon. A simple folding of distribution functions does not take this into account properly. Therefore the contribution from $Q$ values above the $\rho$ mass should be suppressed. A choice of $Q_{\mmax} \approx 1$~GeV is then appropriate for a photoproduction limit description of physics. \end{subentry} \iteme{MSTP(14) :} (D=0) structure of incoming photon beam or target (does not affect photon inside electron, only photons appearing as argument in the {\tt PYINIT} call). \begin{subentry} \iteme{= 0 :} photon is assumed to be point-like, i.e. can only interact in processes which contain an incoming photon, such as ISUB = 33--36. \iteme{= 1 :} photon is assumed to be resolved, i.e. can only interact through its constituent quarks and gluons, in processes such as ISUB = 11--31. Hard processes are calculated with the use of the full photon parton distributions. \iteme{= 2 :} photon is assumed resolved, but only the VMD piece is included in the parton distributions, which therefore mainly is a scaled-down version of the $\rho^0/\pi^0$ one. \iteme {= 3 :} photon is assumed resolved, but only the anomalous part of the photon parton distributions is included. \itemc{Note:} a complete description requires separate runs for the components above, i.e. it is not possible to mix them in a single run. Our best understanding \cite{Sch93,Sch93a} is to have three separate components, $0+2+3$. A simpler alternative is based on two only, $0+1$. \end{subentry} \iteme{MSTP(21) :} (D=1) nature of fermion--fermion scatterings simulated in process 10 by $t$-channel exchange. \begin{subentry} \iteme{= 0 :} all off. \iteme{= 1 :} full mixture of $\gammaZ$ neutral current and $\W^{\pm}$ charged current. \iteme{= 2 :} $\gamma$ neutral current only. \iteme{= 3 :} $\Z^0$ neutral current only. \iteme{= 4 :} $\gammaZ$ neutral current only. \iteme{= 5 :} $\W^{\pm}$ charged current only. \end{subentry} \iteme{MSTP(22) :} (D=0) special override of normal $Q^2$ definition used for maximum of parton-shower evolution, intended for deep inelastic scattering) in lepton--hadron events, see section \ref{ss:showrout}. \iteme{MSTP(23) :} (D=1) for deep inelastic scattering processes (10 and 83), this option allows the $x$ and $Q^2$ of the original hard scattering to be retained by the final electron. \begin{subentry} \iteme{= 0 :} no correction procedure, i.e. $x$ and $Q^2$ of the scattered electron differ from the originally generated $x$ and $Q^2$. \iteme{= 1 :} post facto correction, i.e. the change of electron momentum, by initial and final QCD radiation, primordial $k_{\perp}$ and beam remnant treatment, is corrected for by a shuffling of momentum between the electron and hadron side in the final state. \itemc{Note:} the correction procedure will fail for a fraction of the events, which are thus rejected (and new ones generated in their place). The correction option is not unambiguous, and should not be taken too seriously. For very small $Q^2$ values, the $x$ is not exactly preserved even after this procedure. \end{subentry} \iteme{MSTP(31) :} (D=1) parametrization of total, elastic and diffractive cross sections. \begin{subentry} \iteme{= 0 :} everything is to be set by you yourself in the {\tt PYINT7} common block. For photoproduction, additionally you need to set {\tt VINT(281)}. Normally you would set these values once and for all before the {\tt PYINIT} call, but if you run with variable energies (see {\tt MSTP(171)}) you can also set it before each new {\tt PYEVNT} call. \iteme{= 1 :} Donnachie--Landshoff for total cross section \cite{Don92}, and Schuler--Sj\"ostrand for elastic and diffractive cross sections \cite{Sch93b,Sch93b}. \end{subentry} \iteme{MSTP(32) :} (D=2) $Q^2$ definition in hard scattering for $2 \to 2$ processes. For resonance production $Q^2$ is always chosen to be $\hat{s} = m_R^2$, where $m_R$ is the mass of the resonance. For gauge boson scattering processes $VV \to VV$ the $\W$ or $\Z^0$ squared mass is used as scale in parton distributions. \begin{subentry} \iteme{= 1 :} $Q^2 = 2 \hat{s} \hat{t} \hat{u} / (\hat{s}^2 + \hat{t}^2 + \hat{u}^2)$. \iteme{= 2 :} $Q^2 = (m_{\perp 1}^2 + m_{\perp 2}^2)/2$. \iteme{= 3 :} $Q^2 = \min(-\hat{t}, -\hat{u})$. \iteme{= 4 :} $Q^2 = \hat{s}$. \iteme{= 5 :} $Q^2 = -\hat{t}$. \end{subentry} \iteme{MSTP(33) :} (D=0) inclusion of $K$ factors in hard cross sections for parton--parton interactions (i.e. for incoming quarks and gluons). \begin{subentry} \iteme{= 0 :} none, i.e. $K = 1$. \iteme{= 1 :} a common $K$ factor is used, as stored in {\tt PARP(31)}. \iteme{= 2 :} separate factors are used for ordinary ({\tt PARP(31)}) and colour annihilation graphs ({\tt PARP(32)}). \iteme{= 3 :} A $K$ factor is introduced by a shift in the $\alphas$ $Q^2$ argument, $\alphas = \alphas(${\tt PARP(33)}$Q^2)$. \end{subentry} \iteme{MSTP(34) :} (D=1) use of interference term in matrix elements for QCD processes, see section \ref{sss:QCDjetclass}. \begin{subentry} \iteme{= 0 :} excluded (i.e. string-inspired matrix elements). \iteme{= 1 :} included (i.e. conventional QCD matrix elements). \itemc{Note:} for the option {\tt MSTP(34)=1}, i.e. interference terms included, these terms are divided between the different possible colour configurations according to the pole structure of the (string-inspired) matrix elements for the different colour configurations. \end{subentry} \iteme{MSTP(35) :} (D=0) threshold behaviour for heavy-flavour production, i.e. ISUB = 81, 82, 84, 85, and also for $\Z$ and $\Z'$ decays. The non-standard options are mainly intended for top, but can be used, with less theoretical reliability, also for charm and bottom (for $\Z$ and $\Z'$ only top and heavier flavours are affected). The threshold factors are given in eqs. (\ref{pp:threshenh}) and (\ref{pp:threshsup}). \begin{subentry} \iteme{= 0 :} na\"{\i}ve lowest-order matrix-element behaviour. \iteme{= 1 :} enhancement or suppression close to threshold, according to the colour structure of the process. The $\alphas$ value appearing in the threshold factor (which is not the same as the $\alphas$ of the lowest-order $2 \to 2$ process) is taken to be fixed at the value given in {\tt PARP(35)}. The threshold factor used in an event is stored in {\tt PARI(81)}. \iteme{= 2 :} as {\tt =1}, but the $\alphas$ value appearing in the threshold factor is taken to be running, with argument $Q^2 = m_{\Q} \sqrt{ (\hat{m} - 2m_{\Q})^2 + \Gamma_{\Q}^2}$. Here $m_{\Q}$ is the nominal heavy-quark mass, $\Gamma_{\Q}$ is the width of the heavy-quark-mass distribution, and $\hat{m}$ is the invariant mass of the heavy-quark pair. The $\Gamma_{\Q}$ value has to be stored by the user in {\tt PARP(35)}. The regularization of $\alphas$ at low $Q^2$ is given by {\tt MSTP(36)}. \end{subentry} \iteme{MSTP(36) :} (D=2) regularization of $\alphas$ in the limit $Q^2 \to 0$ for the threshold factor obtainable in the {\tt MSTP(35)=2} option; see {\tt MSTU(115)} for a list of the possibilities. \iteme{MSTP(37) :} (D=1) inclusion of running quark masses in Higgs production ($\q \qbar \to \H^0$) and decay ($\H^0 \to \q \qbar$) couplings. Also included for charged Higgs production and decay, but there only for the down-type quark, since the up-type one normally is a top quark, with $m_{\t} \approx m_{\H}$. \begin{subentry} \iteme{= 0 :} not included, i.e. fixed quark masses are used according to the values in the {\tt PMAS} array. \iteme{= 1 :} included, with running starting from the value given in the {\tt PMAS} array, at a $Q_0$ scale of {\tt PARP(37)} times the quark mass itself, up to a $Q$ scale given by the Higgs mass. \end{subentry} \iteme{MSTP(38) :} (D=5) handling of quark loop masses in the box graphs $\g \g \to \gamma \gamma$ and $\g \g \to \g \gamma$. \begin{subentry} \iteme{= 0 :} the program will for each flavour automatically choose the massless approximation for light quarks and the full massive formulae for heavy quarks, with a dividing line between light and heavy quarks that depends on the actual $\hat{s}$ scale. \iteme{$\geq$1 :} the program will use the massless approximation throughout, assuming the presence of {\tt MSTP(38)} effectively massless quark species (however, at most 8). Normally one would use {\tt =5} for the inclusion of all quarks up to bottom, and {\tt =6} to include top as well. \itemc{Warning:} for {\tt =0}, numerical instabilities may arise for scattering at small angles. Users are therefore recommended in this case to set {\tt CKIN(27)} and {\tt CKIN(28)} so as to exclude the range of scattering angles that are not of interest anyway. \end{subentry} \iteme{MSTP(41) :} (D=1) master switch for all resonance decays ($\Z^0$, $\W^{\pm}$, $\H^0$, $\Z'^0$, $\W'^{\pm}$, $\H'^0$, $\A^0$, $\H^{\pm}$, $\L_{\Q}$, $\R^0$, $\d^*$, $\u^*$, \ldots). \begin{subentry} \iteme{= 0 :} off. \iteme{= 1 :} on. \itemc{Note:} also for {\tt MSTP(41)=1} it is possible to switch off the decays of specific resonances by using the {\tt MDCY(KC,1)} switches in {\Je}. However, since the {\tt MDCY} values are overwritten in the {\tt PYINIT} call, individual resonances should be switched off after the {\tt PYINIT} call. \itemc{Warning:} leptoquark decays must not be switched off if one later on intends to let leptoquarks decay (with {\tt LUEXEC}); see section \ref{sss:LQclass}. \end{subentry} \iteme{MSTP(42) :} (D=1) mass treatment in $2 \to 2$ processes, where the final-state resonances have finite width (see {\tt PARP(41)}). (Does not apply for the production of a single $s$-channel resonance, where the mass appears explicitly in the cross section of the process, and thus is always selected with width.) \begin{subentry} \iteme{= 0 :} particles are put on the mass shell. \iteme{= 1 :} mass generated according to a Breit--Wigner. \end{subentry} \iteme{MSTP(43) :} (D=3) treatment of $\Z^0/\gamma^*$ interference in matrix elements. So far implemented in subprocesses 1, 15, 19, 22, 30 and 35; in other processes what is called a $\Z^0$ is really a $\Z^0$ only, without the $\gamma^*$ piece. \begin{subentry} \iteme{= 1 :} only $\gamma^*$ included. \iteme{= 2 :} only $\Z^0$ included. \iteme{= 3 :} complete $\Z^0/\gamma^*$ structure (with interference) included. \end{subentry} \iteme{MSTP(44) :} (D=7) treatment of $\Z'^0/\Z^0/\gamma^*$ interference in matrix elements. \begin{subentry} \iteme{= 1 :} only $\gamma^*$ included. \iteme{= 2 :} only $\Z^0$ included. \iteme{= 3 :} only $\Z'^0$ included. \iteme{= 4 :} only $\Z^0/\gamma^*$ (with interference) included. \iteme{= 5 :} only $\Z'^0/\gamma^*$ (with interference) included. \iteme{= 6 :} only $\Z'^0/\Z^0$ (with interference) included. \iteme{= 7 :} complete $\Z'^0/\Z^0/\gamma^*$ structure (with interference) included. \end{subentry} \iteme{MSTP(45) :} (D=3) treatment of $\W\W \to \W\W$ structure (ISUB = 77). \begin{subentry} \iteme{= 1 :} only $\W^+\W^+ \to \W^+\W^+$ and $\W^-\W^- \to \W^-\W^-$ included. \iteme{= 2 :} only $\W^+\W^- \to \W^+\W^-$ included. \iteme{= 3 :} all charge combinations $\W\W \to \W\W$ included. \end{subentry} \iteme{MSTP(46) :} (D=1) treatment of $VV \to V'V'$ structures (ISUB = 71--77), where $V$ represents a longitudinal gauge boson. \begin{subentry} \iteme{= 0 :} only $s$-channel Higgs exchange included (where existing). With this option, subprocesses 71--72 and 76--77 will essentially be equivalent to subprocesses 5 and 8, respectively, with the proper decay channels (i.e. only $\Z^0\Z^0$ or $\W^+\W^-$) set via {\tt MDME}. The description obtained for subprocesses 5 and 8 in this case is more sophisticated, however; see section \ref{sss:heavySMHclass}. \iteme{= 1 :} all graphs contributing to $VV \to V'V'$ processes are included. \iteme{= 2 :} only graphs not involving Higgs exchange (either in $s$, $t$ or $u$ channel) are included; this option then gives the na\"{\i}ve behaviour one would expect if no Higgs exists, including unphysical unitarity violations at high energies. \iteme{= 3 :} the strongly interacting Higgs-like model of Dobado, Herrero and Terron \cite{Dob91} with Pad\'e unitarization. Note that to use this option it is necessary to set the Higgs mass to a large number like 20 TeV (i.e. {\tt PMAS(25,1)=20000}). The parameter $\nu$ is stored in {\tt PARP(44)}, but should not have to be changed. \iteme{= 4 :} as {\tt =3}, but with K-matrix unitarization. \iteme{= 5 :} the strongly interacting QCD-like model of Dobado, Herrero and Terron \cite{Dob91} with Pad\'e unitarization. The parameter $\nu$ is stored in {\tt PARP(44)}, but should not have to be changed. The effective techni-$\rho$ mass in this model is stored in {\tt PARP(45)}; by default it is 2054 GeV, which is the expected value for three technicolors, based on scaling up the ordinary $\rho$ mass appropriately. \iteme{= 6 :} as {\tt =5}, but with K-matrix unitarization. While {\tt PARP(45)} still is a parameter of the model, this type of unitarization does not give rise to a resonance at a mass of {\tt PARP(45)}. \end{subentry} \iteme{MSTP(47) :} (D=1) (C) angular orientation of decay products of resonances ($\Z^0$, $\W^{\pm}$, $\H^0$, $\Z'^0$, $\W'^{\pm}$, etc.), either when produced singly or in pairs (also from an $\H^0$ decay), or in combination with a single quark, gluon or photon. \begin{subentry} \iteme{= 0 :} independent decay of each resonance, isotropic in c.m. frame of the resonance. \iteme{= 1 :} correlated decay angular distributions according to proper matrix elements, to the extent these are known. \end{subentry} \iteme{MSTP(48) :} (D=2) (C) possibility to switch between top decay before or after fragmentation. As a rule of thumb, option 0 is recommendable for top masses below 120 GeV and option 1 above that, but clearly there is a gradual transition between the two. \begin{subentry} \iteme{= 0 :} top quarks fragment to top hadrons, which subsequently decay. The $\b$ quark may be allowed to shower, see {\tt MSTJ(27)}. The $\W$ produced in the decay is code 89. \iteme{= 1 :} top quarks decay, $t \to \b \W^+$, and thereafter the $\b$ quark fragments. Parton showering of the $\b$ is automatically included, but can be switched off with {\tt MSTP(71)}. The $\W$ has the ordinary code 24, and is allowed to decay isotropically. \iteme{= 2 :} as {\tt =1}, except that the $\W$ decay is anisotropic, as expected from $\W$ polarization in the top decay. \itemc{Note:} in neither case is top treated as an ordinary resonance, i.e. the cross sections are not automatically reduced if only some decay channels are left open. \end{subentry} \iteme{MSTP(49) :} (D=2) (C) possibility to switch between fourth generation decay before or after fragmentation. For the quarks $\hrm$ and $\lrm$ the meaning is exactly as {\tt MSTP(48)} is for the $\t$ quark. For the lepton $\chi$ the difference is whether decay is handled as part of the {\Py} resonance machinery or as part of the {\Je} particle decay one. The $\nu_{\chi}$ is assumed stable, so the option above would currently make no difference. \begin{subentry} \iteme{= 0 :} hadrons are first produced, which subsequently decay (in {\tt LUDECY}). The new quark may be allowed to shower, see {\tt MSTJ(27)}. The $\W$ produced in the decay is code $\pm89$. \iteme{= 1 :} the heavy quark first decays (in {\tt PYRESD}) to a light one, and thereafter the light quark fragments. Parton showering in the decay is automatically included, but can be switched off with {\tt MSTP(71)}. The $\W$ has the ordinary code $\pm24$, and is allowed to decay isotropically. \iteme{= 2 :} as {\tt =1}, except that the $\W$ decay is anisotropic, as expected from $\W$ polarization in the heavy flavour decay. \itemc{Note:} in neither case is fermion treated as an ordinary resonance, i.e. the cross sections are not automatically reduced if only some decay channels are left open. \end{subentry} \iteme{MSTP(51) :} (D=9) choice of proton parton-distribution set; see also {\tt MSTP(52)}. \begin{subentry} \iteme{= 1 :} EHLQ set 1 (1986 updated version). \iteme{= 2 :} EHLQ set 2 (1986 updated version). \iteme{= 3 :} Duke--Owens set 1. \iteme{= 4 :} Duke--Owens set 2. \iteme{= 5 :} CTEQ2M (best $\br{\mrm{MS}}$ fit). \iteme{= 6 :} CTEQ2MS (singular at small $x$). \iteme{= 7 :} CTEQ2MF (flat at small $x$). \iteme{= 8 :} CTEQ2ML (large $\Lambda$). \iteme{= 9 :} CTEQ2L (best leading order fit). \iteme{= 10 :} CTEQ2D (best DIS fit). \itemc{Note:} since all parametrizations have some region of applicability, the parton distributions are assumed frozen below the lowest $Q^2$ covered by the parametrizations; the CTEQ2 ones have been allowed to extend down to $Q_{\mmin} = 1$~GeV. For the former four, evolution is also frozed above the maximum $Q^2$. The extrapolation of EHLQ to low $x$ is covered by {\tt PARP(51)}. \end{subentry} \iteme{MSTP(52) :} (D=1) choice of proton parton-distribution-function library. \begin{subentry} \iteme{= 1 :} the internal {\Py} one, with parton distributions according to the {\tt MSTP(51)} above. \iteme{= 2 :} the {\sc Pdflib} one \cite{Plo93}, with the {\sc Pdflib} (version 4) {\tt NGROUP} and {\tt NSET} numbers to be given as {\tt MSTP(51) = 1000 * NGROUP + NSET}. \itemc{Note:} to make use of option 2, it is necessary to link {\sc Pdflib}. Additionally, on most computers, the two dummy routines {\tt PDFSET} and {\tt STRUCTM} at the end of the {\Py} file should be removed or commented out. \itemc{Warning:} For external parton distribution libraries, {\Py} does not check whether {\tt MSTP(51)} corresponds to a valid code, or if special $x$ and $Q^2$ restrictions exist for a given set, such that crazy values could be returned. This puts an extra responsibility on you. \end{subentry} \iteme{MSTP(53) :} (D=1) choice of pion parton-distribution set; see also {\tt MSTP(54)}. \begin{subentry} \iteme{= 1 :} Owens set 1. \iteme{= 2 :} Owens set 2. \end{subentry} \iteme{MSTP(54) :} (D=1) choice of pion parton-distribution-function library. \begin{subentry} \iteme{= 1 :} the internal {\Py} one, with parton distributions according to the {\tt MSTP(53)} above. \iteme{= 2 :} the {\sc Pdflib} one \cite{Plo93}, with the {\sc Pdflib} (version 4) {\tt NGROUP} and {\tt NSET} numbers to be given as {\tt MSTP(53) = 1000 * NGROUP + NSET}. \itemc{Note:} to make use of option 2, it is necessary to link {\sc Pdflib}. Additionally, on most computers, the two dummy routines {\tt PDFSET} and {\tt STRUCTM} at the end of the {\Py} file should be removed or commented out. \itemc{Warning:} For external parton distribution libraries, {\Py} does not check whether {\tt MSTP(53)} corresponds to a valid code, or if special $x$ and $Q^2$ restrictions exist for a given set, such that crazy values could be returned. This puts an extra responsibility on you. \end{subentry} \iteme{MSTP(55) :} (D=1) choice of photon parton-distribution set; see also {\tt MSTP(56)}. Currently there are no alternatives available for this switch (unless {\tt MSTP(56)} is changed). \begin{subentry} \iteme{= 1 :} Drees--Grassie. \end{subentry} \iteme{MSTP(56) :} (D=1) choice of photon parton-distribution-function library. \begin{subentry} \iteme{= 1 :} the internal {\Py} one, with parton distributions according to the {\tt MSTP(55)} above. \iteme{= 2 :} the {\sc Pdflib} one \cite{Plo93}, with the {\sc Pdflib} (version 4) {\tt NGROUP} and {\tt NSET} numbers to be given as {\tt MSTP(55) = 1000 * NGROUP + NSET}. \itemc{Note:} to make use of option 2, it is necessary to link {\sc Pdflib}. Additionally, on most computers, the two dummy routines {\tt PDFSET} and {\tt STRUCTM} at the end of the {\Py} file should be removed or commented out. \itemc{Warning:} For external parton-distribution libraries, {\Py} does not check whether {\tt MSTP(55)} corresponds to a valid code, or if special $x$ and $Q^2$ restrictions exist for a given set, such that crazy values could be returned. This puts an extra responsibility on you. \end{subentry} \iteme{MSTP(57) :} (D=1) choice of $Q^2$ dependence in parton-distribution functions. \begin{subentry} \iteme{= 0 :} parton distributions are evaluated at nominal lower cut-off value $Q_0^2$, i.e. are made $Q^2$-independent. \iteme{= 1 :} the parametrized $Q^2$ dependence is used. \iteme{= 2 :} the parametrized parton-distribution behaviour is kept at large $Q^2$ and $x$, but modified at small $Q^2$ and/or $x$, so that parton distributions vanish in the limit $Q^2 \to 0$ and have a theoretically motivated small-$x$ shape \cite{Sch93a}. This option is only valid for the $\p$ and $\n$. \end{subentry} \iteme{MSTP(58) :} (D=min(6, 2$\times${\tt MSTP(1)})) maximum number of quark flavours used in parton distributions, and thus also for initial-state space-like showers. If some distributions (notably $\t$) are absent in the parametrization selected in {\tt MSTP(51)}, these are obviously automatically excluded. \iteme{MSTP(61) :} (D=1) (C) master switch for initial-state QCD and QED radiation. \begin{subentry} \iteme{= 0 :} off. \iteme{= 1 :} on. \end{subentry} \iteme{MSTP(62) - MSTP(65) :} (C) further switches for initial-state radiation, see section \ref{ss:showrout}. \iteme{MSTP(71) :} (D=1) (C) master switch for final-state QCD and QED radiation. \begin{subentry} \iteme{= 0 :} off. \iteme{= 1 :} on. \itemc{Note:} additional switches (e.g. for conventional/coherent showers) are available in {\tt MSTJ(41) - MSTJ(50)} and {\tt PARJ(81) - PARJ(88)}, see section \ref{ss:showrout}. \end{subentry} \iteme{MSTP(81) :} (D=1) master switch for multiple interactions. \begin{subentry} \iteme{= 0 :} off. \iteme{= 1 :} on. \end{subentry} \iteme{MSTP(82) - MSTP(83) :} further switches for multiple interactions, see section \ref{ss:multintpar}. \iteme{MSTP(85) :} possibility to dampen hard scattering matrix elements in the limit $\pT \to 0$. It parellels some of the multiple interactions options for QCD processes, but can be used for any $2 \to 2$ process. \begin{subentry} \iteme{= 0 :} off, i.e. standard matrix elements are kept. \iteme{= 1 :} on, i.e. matrix elements are multiplied by a factor $\pT^4/(\pT^2 + p_{\perp 0}^2)^2$, where $p_{\perp 0}$ is given by {\tt PARP(82)}. Additionally $\alphas$ is evaluated at a scale $\pT^2 + p_{\perp 0}^2$ rather than just $\pT^2$. \end{subentry} \iteme{MSTP(91) - MSTP(93) :} switches for beam remnant treatment, see section \ref{ss:multintpar}. \iteme{MSTP(101) :} (D=3) (C) structure of diffractive system. \begin{subentry} \iteme{= 1 :} forward moving diquark + interacting quark. \iteme{= 2 :} forward moving diquark + quark joined via interacting gluon (`hairpin' configuration). \iteme{= 3 :} a mixture of the two options above, with a fraction {\tt PARP(101)} of the former type. \end{subentry} \iteme{MSTP(102) :} (D=1) (C) decay of a $\rho^0$ meson produced by `elastic' scattering of an incoming $\gamma$, as in $\gamma \p \to \rho^0 \p$, or the same with the hadron diffractively excited. \begin{subentry} \iteme{= 0 :} the $\rho^0$ is allowed to decay isotropically, like any other $\rho^0$. \iteme{= 1 :} the decay $\rho^0 \to \pi^+ \pi^-$ is done with an angular distribution proportional to $\sin^2 \theta$ in its rest frame, where the $z$ axis is given by the direction of motion of the $\rho^0$. The $\rho^0$ decay is then done as part of the hard process, i.e. also when {\tt MSTP(111)=0}. \end{subentry} \iteme{MSTP(111) :} (D=1) (C) master switch for fragmentation and decay, as obtained with a {\tt LUEXEC} call. \begin{subentry} \iteme{= 0 :} off. \iteme{= 1 :} on. \iteme{= -1 :} only choose kinematical variables for hard scattering, i.e. no jets are defined. This is useful, for instance, to calculate cross sections (by Monte Carlo integration) without wanting to simulate events; information obtained with {\tt PYSTAT(1)} will be correct. \end{subentry} \iteme{MSTP(112) :} (D=1) (C) cuts on partonic events; only affects an exceedingly tiny fraction of events. \begin{subentry} \iteme{= 0 :} no cuts (can be used only with independent fragmentation, at least in principle). \iteme{= 1 :} string cuts (as normally required for fragmentation). \end{subentry} \iteme{MSTP(113) :} (D=1) (C) recalculation of energies of partons from their momenta and masses, to be done immediately before and after fragmentation, to partly compensate for some numerical problems appearing at high energies. \begin{subentry} \iteme{= 0 :} not performed. \iteme{= 1 :} performed. \end{subentry} \iteme{MSTP(121) :} (D=0) calculation of kinematics selection coefficients and differential cross section maxima for included (by user or default) subprocesses. \begin{subentry} \iteme{= 0 :} not known; to be calculated at initialization. \iteme{= 1 :} not known; to be calculated at initialization; however, the maximum value then obtained is to be multiplied by {\tt PARP(121)} (this may be useful if a violation factor has been observed in a previous run of the same kind). \iteme{= 2 :} known; kinematics selection coefficients stored by user in {\tt COEF(ISUB,J)} ({\tt J} = 1--20) in common block {\tt PYINT2} and maximum of the corresponding differential cross section times Jacobians in {\tt XSEC(ISUB,1)} in common block {\tt PYINT5}. This is to be done for each included subprocess ISUB before initialization, with the sum of all {\tt XSEC(ISUB,1)} values, except for ISUB = 95, stored in {\tt XSEC(0,1)}. \end{subentry} \iteme{MSTP(122) :} (D=1) initialization and differential cross section maximization print-out. \begin{subentry} \iteme{= 0 :} none. \iteme{= 1 :} short message. \iteme{= 2 :} detailed message, including full maximization. \end{subentry} \iteme{MSTP(123) :} (D=2) reaction to violation of maximum differential cross section. \begin{subentry} \iteme{= 0 :} stop generation, print message. \iteme{= 1 :} continue generation, print message for each subsequently larger violation. \iteme{= 2 :} as {\tt =1}, but also increase value of maximum. \end{subentry} \iteme{MSTP(124) :} (D=1) (C) frame for presentation of event. \begin{subentry} \iteme{= 1 :} as specified in {\tt PYINIT}. \iteme{= 2 :} c.m. frame of incoming particles. \end{subentry} \iteme{MSTP(125) :} (D=1) (C) documentation of partonic process, see section \ref{sss:PYrecord} for details. \begin{subentry} \iteme{= 0 :} only list ultimate string/particle configuration. \iteme{= 1 :} additionally list short summary of the hard process. \iteme{= 2 :} list complete documentation of intermediate steps of parton-shower evolution. \end{subentry} \iteme{MSTP(126) :} (D=20) number of lines at the beginning of event record that are reserved for event-history information; see section \ref{sss:PYrecord}. This value should never be reduced, but may be increased at a later date if more complicated processes are included. \iteme{MSTP(128) :} (D=0) storing of copy of resonance decay products in the documentation section of the event record, and mother pointer ({\tt K(I,3)}) relation of the actual resonance decay products (stored in the main section of the event record) to the documentation copy. \begin{subentry} \iteme{= 0 :} products are stored also in the documentation section, and each product stored in the main section points back to the corresponding entry in the documentation section. \iteme{= 1 :} products are stored also in the documentation section, but the products stored in the main section point back to the decaying resonance copy in the main section. \iteme{= 2 :} products are not stored in the documentation section; the products stored in the main section point back to the decaying resonance copy in the main section. \end{subentry} \iteme{MSTP(129) :} (D=10) for the maximization of $2 \to 3$ processes ({\tt ISET(ISUB)=5}) each phase-space point in $\tau$, $y$ and $\tau'$ is tested {\tt MSTP(129)} times in the other dimensions (at randomly selected points) to determine the effective maximum in the ($\tau$, $y$, $\tau'$) point. \iteme{MSTP(131) :} (D=0) master switch for pile-up events, i.e. several independent hadron--hadron interactions generated in the same bunch--bunch crossing, with the events following one after the other in the event record. \begin{subentry} \iteme{= 0 :} off, i.e. only one event is generated at a time. \iteme{= 1 :} on, i.e. several events are allowed in the same event record. Information on the processes generated may be found in {\tt MSTI(41) - MSTI(50)}. \end{subentry} \iteme{MSTP(132) - MSTP(134) :} further switches for pile-up events, see section \ref{ss:multintpar}. \iteme{MSTP(141) :} (D=0) calling of {\tt PYKCUT} in the event-generation chain, for inclusion of user-specified cuts. \begin{subentry} \iteme{= 0 :} not called. \iteme{= 1 :} called. \end{subentry} \iteme{MSTP(142) :} (D=0) calling of {\tt PYEVWT} in the event-generation chain, either to give weighted events or to modify standard cross sections. See {\tt PYEVWT} description in section \ref{ss:PYTmainroutines} for further details. \begin{subentry} \iteme{= 0 :} not called. \iteme{= 1 :} called; the distribution of events among subprocesses and in kinematics variables is modified by the factor {\tt WTXS}, set by the user in the {\tt PYEVWT} call, but events come with a compensating weight {\tt PARI(10)=1./WTXS}, such that total cross sections are unchanged. \iteme{= 2 :} called; the cross section itself is modified by the factor {\tt WTXS}, set by the user in the {\tt PYEVWT} call. \end{subentry} \iteme{MSTP(151) :} (D=0) introduce smeared position of primary vertex of events. \begin{subentry} \iteme{= 0 :} no, i.e. the primary vertex of each event is at the origin. \iteme{= 1 :} yes, with Gaussian distributions separately in $x$, $y$, $z$ and $t$. The respective widths of the Gaussians have to be given in {\tt PARP(151) - PARP(154)}. Also pile-up events obtain separate primary vertices. No provisions are made for more complicated beam-spot shapes, e.g. with a spread in $z$ that varies as a function of $t$. Note that a large beam spot combined with some of the {\tt MSTJ(22)} options may lead to many particles not being allowed to decay at all. \end{subentry} \iteme{MSTP(171) :} (D=0) possibility of variable energies from one event to the next. For further details see subsection \ref{ss:PYvaren}. \begin{subentry} \iteme{= 0 :} no; i.e. the energy is fixed at the initialization call. \iteme{= 1 :} yes; i.e. a new energy has to be given for each new event. \end{subentry} \iteme{MSTP(172) :} (D=2) options for generation of events with variable energies, applicable when {\tt MSTP(171)=1}. \begin{subentry} \iteme{= 1 :} an event is generated at the requested energy, i.e. internally a loop is performed over possible event configurations until one is accepted. If the requested c.m. energy of an event is below {\tt PARP(2)} the run is aborted. Cross-section information can not be trusted with this option, since it depends on how you decided to pick the requested energies. \iteme{= 2 :} only one event configuration is tried. If that is accepted, the event is generated in full. If not, no event is generated, and the status code {\tt MSTI(61)=1} is returned. You are then expected to give a new energy, looping until an acceptable event is found. No event is generated if the requested c.m. energy is below {\tt PARP(2)}, instead {\tt MSTI(61)=1} is set to signal the failure. In principle, cross sections should come out correctly with this option. \end{subentry} \iteme{MSTP(173) :} (D=0) possibility for user to give in an event weight to compensate for a biased choice of beam spectrum. \begin{subentry} \iteme{= 0 :} no, i.e. event weight is unity. \iteme{= 1 :} yes; weight to be given for each event in {\tt PARP(173)}. \end{subentry} \iteme{MSTP(181) :} (R) {\Py} version number. \iteme{MSTP(182) :} (R) {\Py} subversion number. \iteme{MSTP(183) :} (R) last year of change for {\Py}. \iteme{MSTP(184) :} (R) last month of change for {\Py}. \iteme{MSTP(185) :} (R) last day of change for {\Py}. \iteme{MSTP(186) :} (R) earliest subversion of {\Je} version 7 with which this {\Py} subversion can be run. \boxsep \iteme{PARP(1) :}\label{p:PARP} (D=0.25 GeV) nominal $\Lambda_{\mrm{QCD}}$ used in running $\alphas$ for hard scattering (see {\tt MSTP(3)}). \iteme{PARP(2) :} (D=10. GeV) lowest c.m. energy for the event as a whole that the program will accept to simulate. \iteme{PARP(13) :} (D=1. GeV$^2$) $Q_{\mmax}^2$ scale, to be set by user for defining maximum scale allowed for photoproduction when using the option {\tt MSTP(13)=2}. \iteme{PARP(14) :} (D=0.01) in the numerical integration of quark and gluon parton distributions inside an electron, the successive halvings of evaluation-point spacing is interrupted when two values agree in relative size, $|$new$-$old$|$/(new$+$old), to better than {\tt PARP(14)}. There are hardwired lower and upper limits of 2 and 8 halvings, respectively. \iteme{PARP(15) :} (D=0.5 GeV) lower cut-off $p_0$ used to define minimum transverse momentum in branchings $\gamma \to \q\qbar$ in the anomalous event class of $\gamma\p$ interactions. \iteme{PARP(16) :} (D=1.) the anomalous parton-distribution functions of the photon are taken to have the charm and bottom flavour thresholds at virtuality {\tt PARP(16)}$\times m_{\q}^2$. \iteme{PARP(31) :} (D=1.5) common $K$ factor multiplying the differential cross section for hard parton--parton processes when {\tt MSTP(33)=1} or {\tt 2}, with the exception of colour annihilation graphs in the latter case. \iteme{PARP(32) :} (D=2.0) special $K$ factor multiplying the differential cross section in hard colour annihilation graphs, including resonance production, when {\tt MSTP(33)=2}. \iteme{PARP(33) :} (D=0.075) this factor is used to multiply the ordinary $Q^2$ scale in $\alphas$ at the hard interaction for {\tt MSTP(33)=3}. The effective $K$ factor thus obtained is in accordance with the results in \cite{Ell86}. \iteme{PARP(35) :} (D=0.20) fix $\alphas$ value that is used in the heavy-flavour threshold factor when {\tt MSTP(35)=1}. \iteme{PARP(36) :} (D=0. GeV) the width $\Gamma_{\Q}$ for the heavy flavour studied in processes ISUB = 81 or 82; to be used for the threshold factor when {\tt MSTP(35)=2}. \iteme{PARP(37) :} (D=2.) for {\tt MSTP(37)=1} this regulates the point at which the reference on-shell quark mass in Higgs couplings is assumed defined; specifically the running quark mass is assumed to coincide with the fix one at an energy scale {\tt PARP(37)} times the fix quark mass, i.e. $m_{\mrm{running}}(${\tt PARP(37)}$\times m_{\mrm{fix}}) = m_{\mrm{fix}}$. \iteme{PARP(38) :} (D=0.70 GeV$^3$) the squared wave function at the origin, $|R(0)|^2$, of the $\Jpsi$ wave function. Used for process 86. See ref. \cite{Glo88}. \iteme{PARP(39) :} (D=0.006 GeV$^3$) the squared derivative of the wave function at the origin, $|R'(0)|^2/m^2$, of the $\chi_{\c}$ wave functions. Used for the processes 87, 88 and 89. See ref. $\cite{Glo88}$. \iteme{PARP(41) :} (D=0.020 GeV) in the process of generating mass for resonances, and optionally to force that mass to be in a given range, only resonances with a total width in excess of {\tt PARP(41)} are generated according to a Breit--Wigner shape (if allowed by {\tt MSTP(42)}), while narrower resonances are put on the mass shell. \iteme{PARP(42) :} (D=2. GeV) minimum mass of resonances assumed to be allowed when evaluating total width of $\H^0$ to $\Z^0 \Z^0$ or $\W^+ \W^-$ for cases when the $\H^0$ is so light that (at least) one $\Z/\W$ is forced to be off the mass shell. Also generally used as safety check on minimum mass of resonance. Note that some {\tt CKIN} values may provide additional constraints. \iteme{PARP(43) :} (D=0.10) precision parameter used in numerical integration of width into channel with at least one daughter off the mass shell. \iteme{PARP(44) :} (D=1000.) the $\nu$ parameter of the strongly interacting $\Z/\W$ model of Dobado, Herrero and Terron \cite{Dob91}. \iteme{PARP(45) :} (D=2054. GeV) the effective techni-$\rho$ mass parameter of the strongly interacting model of Dobado, Herrero and Terron \cite{Dob91}; see {\tt MSTP(46)=5}. On physical grounds it should not be chosen smaller than about 1 TeV or larger than about the default value. \iteme{PARP(46) :} (D=123. GeV) the $F_{\pi}$ decay constant that appears inversely quadratically in all techni-$\eta$ partial decay widths \cite{Eic84,App92}. \iteme{PARP(47) :} (D=246. GeV) vacuum expectation value $v$ used in the DHT scenario \cite{Dob91} to define the width of the techni-$\rho$; this width is inversely proportional $v^2$. \iteme{PARP(51) :} (D=1.) if parton distributions for light flavours have to be extrapolated to $x$ values lower than covered by the parametrizations, an $x^{-b}$ behaviour, with $b =${\tt PARP(51)}, is assumed in that region. This option only applies for the EHLQ proton parton distributions that are internal to {\Py}. \iteme{PARP(61) - PARP(65) :} (C) parameters for initial-state radiation, see section \ref{ss:showrout}. \iteme{PARP(71) :} (C) parameter for final-state radiation, see section \ref{ss:showrout}. \iteme{PARP(81) - PARP(88) :} parameters for multiple interactions, see section \ref{ss:multintpar}. \iteme{PARP(91) - PARP(100) :} parameters for beam remnant treatment, see section \ref{ss:multintpar}. \iteme{PARP(101) :} (D=0.50) fraction of diffractive systems in which a quark is assumed kicked out by the pomeron rather than a gluon; applicable for option {\tt MSTP(101)=3}. \iteme{PARP(102) :} (D=0.28 GeV) the mass spectrum of diffractive states (in single and double diffractive scattering) is assumed to start {\tt PARP(102)} above the mass of the particle that is diffractively excited. In this connection, an incoming $\gamma$ is taken to have the selected VMD meson mass, i.e. $m_{\rho}$, $m_{\omega}$, $m_{\phi}$ or $m_{\Jpsi}$. \iteme{PARP(103) :} (D=1.0 GeV) if the mass of a diffractive state is less than {\tt PARP(103)} above the mass of the particle that is diffractively excited, the state is forced to decay isotropically into a two-body channel. In this connection, an incoming $\gamma$ is taken to have the selected VMD meson mass, i.e. $m_{\rho}$, $m_{\omega}$, $m_{\phi}$ or $m_{\Jpsi}$. If the mass is higher than this threshold, the standard string fragmentation machinery is used. The forced two-body decay is always carried out, also when {\tt MSTP(111)=0}. \iteme{PARP(111) :} (D=2. GeV) used to define the minimum invariant mass of the remnant hadronic system (i.e. when interacting partons have been taken away), together with original hadron masses and extra parton masses. \iteme{PARP(121) :} (D=1.) the maxima obtained at initial maximization are multiplied by this factor if {\tt MSTP(121)=1}; typically {\tt PARP(121)} would be given as the product of the violation factors observed (i.e. the ratio of final maximum value to initial maximum value) for the given process(es). \iteme{PARP(122) :} (D=0.4) fraction of total probability that is shared democratically between the {\tt COEF} coefficients open for the given variable, with the remaining fraction distributed according to the optimization results of {\tt PYMAXI}. \iteme{PARP(131) :} parameter for pile-up events, see section \ref{ss:multintpar}. \iteme{PARP(151) - PARP(154) :} (D=4*0.) (C) regulate the assumed beam-spot size. For {\tt MSTP(151)=1} the $x$, $y$, $z$ and $t$ coordinates of the primary vertex of each event are selected according to four independent Gaussians. The widths of these Gaussians are given by the four parameters, where the first three are in units of mm and the fourth in mm/$c$. \iteme{PARP(161) - PARP(164) :} (D=2.20, 23.6, 18.4, 11.5) couplings $f_V^2/4\pi$ of the photon to the $\rho^0$, $\omega$, $\phi$ and $\Jpsi$ vector mesons. \iteme{PARP(171) :} to be set, event-by-event, when variable energies are allowed, i.e. when {\tt MSTP(171)=1}. If {\tt PYINIT} is called with {\tt FRAME='CMS'} ({\tt ='FIXT'}), {\tt PARP(171)} multiplies the c.m. energy (beam energy) used at initialization. For the options {\tt 'USER'}, {\tt 'FOUR'} and {\tt 'FIVE'}, {\tt PARP(171)} is dummy, since there the momenta are set in the {\tt P} array. \iteme{PARP(173) :} event weight to be given by user when {\tt MSTP(173)=1}. \end{entry} \subsection{General Event Information} When an event is generated with {\tt PYEVNT}, some information on it is stored in the {\tt MSTI} and {\tt PARI} arrays of the {\tt PYPARS} common block (often copied directly from the internal {\tt MINT} and {\tt VINT} variables). Further information is stored in the complete event record; see section \ref{ss:evrec}. Part of the information is only relevant for some subprocesses; by default everything irrelevant is set to 0. Kindly note that, like the {\tt CKIN} constraints described in section \ref{ss:PYswitchkin}, kinematical variables normally (i.e. where it is not explicitly stated otherwise) refer to the na\"{\i}ve hard scattering, before initial- and final-state radiation effects have been included. \drawbox{COMMON/PYPARS/MSTP(200),PARP(200),MSTI(200),PARI(200)}% \label{p:PYPARS2} \begin{entry} \itemc{Purpose:} to provide information on latest event generated or, in a few cases, on statistics accumulated during the run. \iteme{MSTI(1) :}\label{p:MSTI} specifies the general type of subprocess that has occurred, according to the ISUB code given in section \ref{ss:ISUBcode}. \iteme{MSTI(2) :} whenever {\tt MSTI(1)} (together with {\tt MSTI(15)} and {\tt MSTI(16)}) are not enough to specify the type of process uniquely, {\tt MSTI(2)} provides an ordering of the different possibilities. This is particularly relevant for the different colour-flow topologies possible in QCD $2 \to 2$ processes. With $i = ${\tt MSTI(15)}, $j = ${\tt MSTI(16)} and $k = ${\tt MSTI(2)}, the QCD possibilities are, in the classification scheme of \cite{Ben84} (cf. section \ref{sss:QCDjetclass}): \begin{subentry} \itemn{ISUB = 11,} $i = j$, $\q_i \q_i \to \q_i \q_i$; \\ $k = 1$ : colour configuration $A$. \\ $k = 2$ : colour configuration $B$. \itemn{ISUB = 11,} $i \neq j$, $\q_i \q_j \to \q_i \q_j$; \\ $k = 1$ : only possibility. \itemn{ISUB = 12,} $\q_i \qbar_i \to \q_l \qbar_l$; \\ $k = 1$ : only possibility. \itemn{ISUB = 13,} $\q_i \qbar_i \to \g \g$; \\ $k = 1$ : colour configuration $A$. \\ $k = 2$ : colour configuration $B$. \itemn{ISUB = 28,} $\q_i \g \to \q_i \g$; \\ $k = 1$ : colour configuration $A$. \\ $k = 2$ : colour configuration $B$. \itemn{ISUB = 53,} $\g \g \to \q_l \qbar_l$; \\ $k = 1$ : colour configuration $A$. \\ $k = 2$ : colour configuration $B$. \itemn{ISUB = 68,} $\g \g \to \g \g$; \\ $k = 1$ : colour configuration $A$. \\ $k = 2$ : colour configuration $B$. \\ $k = 3$ : colour configuration $C$. \itemn{ISUB = 83,} $\f \q \to \f' \Q$ (by $t$-channel $\W$ exchange; does not distinguish colour flows but result of user selection); \\ $k = 1$ : heavy flavour $\Q$ is produced on side 1. \\ $k = 2$ : heavy flavour $\Q$ is produced on side 2. \end{subentry} \iteme{MSTI(3) :} the number of partons produced in the hard interactions, i.e. the number $n$ of the $2 \to n$ matrix elements used; it is sometimes 3 or 4 when a basic $2 \to 1$ or $2 \to 2$ process has been folded with two $1 \to 2$ initial branchings (like $\q \q' \to \q'' \q''' \H^0$). \iteme{MSTI(4) :} number of documentation lines at the beginning of the common block {\tt LUJETS} that are given with {\tt K(I,1)=21}; 0 for {\tt MSTP(125)=0}. \iteme{MSTI(5) :} number of events generated to date in current run. \iteme{MSTI(6) :} current frame of event, cf. {\tt MSTP(124)}. \iteme{MSTI(7), MSTI(8) :} line number for documentation of outgoing partons/particles from hard scattering for $2 \to 2$ or $2 \to 1 \to 2$ processes (else = 0). \iteme{MSTI(10) :} is 1 if cross section maximum was violated in current event, and 0 if not. \iteme{MSTI(11) :} KF flavour code for beam (side 1) particle. \iteme{MSTI(12) :} KF flavour code for target (side 2) particle. \iteme{MSTI(13), MSTI(14) :} KF flavour codes for side 1 and side 2 initial-state shower initiators. \iteme{MSTI(15), MSTI(16) :} KF flavour codes for side 1 and side 2 incoming partons to the hard interaction. \iteme{MSTI(17), MSTI(18) :} flag to signal if particle on side 1 or side 2 has been scattered diffractively; 0 if no, 1 if yes. \iteme{MSTI(21) - MSTI(24) :} KF flavour codes for outgoing partons from the hard interaction. The number of positions actually used is process-dependent, see {\tt MSTI(3)}; trailing positions not used are set = 0. \iteme{MSTI(25), MSTI(26) :} KF flavour codes of the products in the decay of a single $s$-channel resonance formed in the hard interaction. Are thus only used when {\tt MSTI(3)=1} and the resonance is allowed to decay. \iteme{MSTI(31) :} number of hard or semi-hard scatterings that occurred in the current event in the multiple-interaction scenario; is = 0 for a low-$\pT$ event. \iteme{MSTI(41) :} the number of pile-up events generated in the latest {\tt PYEVNT} call (including the first, `hard' event). \iteme{MSTI(42) - MSTI(50) :} ISUB codes for the events 2--10 generated in the pile-up-events scenario. The first event ISUB code is stored in {\tt MSTI(1)}. If {\tt MSTI(41)} is less than 10, only as many positions are filled as there are pile-up events. If MSTI(41) is above 10, some ISUB codes will not appear anywhere. \iteme{MSTI(51) :} normally 0 but set to 1 if a {\tt PYUPEV} call did not return an event, such that {\tt PYEVNT} could not generate an event. For further details, see end of section \ref{ss:PYnewproc}. \iteme{MSTI(52) :} counter for the number of times the current event configuration failed in the generation machinery. For accepted events this is always 0, but the counter can be used inside {\tt PYUPEV} to check on anomalous occurrences. For further details, see end of section \ref{ss:PYnewproc}. \iteme{MSTI(61) :} status flag set when events are generated. It is only of interest for runs with variable energies, {\tt MSTP(171)=1}, with the option {\tt MSTP(172)=2}. \begin{subentry} \iteme{= 0 :} an event has been generated. \iteme{= 1 :} no event was generated, either because the c.m. energy was too low or because the Monte Carlo phase space point selection machinery rejected the trial point. A new energy is to be picked by the user. \end{subentry} \boxsep \iteme{PARI(1) :}\label{p:PARI} total integrated cross section for the processes under study, in mb. This number is obtained as a by-product of the selection of hard-process kinematics, and is thus known with better accuracy when more events have been generated. The value stored here is based on all events until the latest one generated. \iteme{PARI(2) :} is the ratio {\tt PARI(1)/MSTI(5)}, i.e. the ratio of total integrated cross section and number of events generated. Histograms filled with unit event weight have to be multiplied by this factor, at the end of the run, to convert results to mb. For {\tt MSTP(142)=1}, {\tt MSTI(5)} is replaced by the sum of {\tt PARI(10)} values. Histograms are then filled with weight {\tt PARI(10)} for each event and multiplied by {\tt PARI(2)} at the end. \iteme{PARI(9) :} is weight {\tt WTXS} returned from {\tt PYEVWT} call when {\tt MSTP(142)}$\geq 1$, otherwise is 1. \iteme{PARI(10) :} is compensating weight {\tt 1./WTXS} that should be associated to events when {\tt MSTP(142)=1}, else is 1. \iteme{PARI(11) :} $E_{\mrm{cm}}$, i.e. total c.m. energy. \iteme{PARI(12) :} $s$, i.e. squared total c.m. energy. \iteme{PARI(13) :} $\hat{m} = \sqrt{\hat{s}}$, i.e. mass of the hard-scattering subsystem. \iteme{PARI(14) :} $\hat{s}$ of the hard subprocess ($2 \to 2$ or $2 \to 1$). \iteme{PARI(15) :} $\hat{t}$ of the hard subprocess ($2 \to 2$ or $2 \to 1 \to 2$). \iteme{PARI(16) :} $\hat{u}$ of the hard subprocess ($2 \to 2$ or $2 \to 1 \to 2$). \iteme{PARI(17) :} $\hat{p}_{\perp}$ of the hard subprocess ($2 \to 2$ or $2 \to 1 \to 2$), evaluated in the rest frame of the hard interaction. \iteme{PARI(18) :} $\hat{p}_{\perp}^2$ of the hard subprocess; see {\tt PARI(17)}. \iteme{PARI(19) :} $\hat{m}'$, the mass of the complete three- or four-body final state in $2 \to 3$ or $2 \to 4$ processes (while $\hat{m}$, given in {\tt PARI(13)}, here corresponds to the one- or two-body central system). Kinematically $\hat{m} \leq \hat{m}' \leq E_{\mrm{cm}}$. \iteme{PARI(20) :} $\hat{s}' = \hat{m}'^2$; see {\tt PARI(19)}. \iteme{PARI(21) :} $Q$ of the hard-scattering subprocess. The exact definition is process-dependent, see {\tt MSTP(32)}. \iteme{PARI(22) :} $Q^2$ of the hard-scattering subprocess; see {\tt PARI(21)}. \iteme{PARI(23) :} $Q$ of the outer hard-scattering subprocess. Agrees with {\tt PARI(21)} for a $2 \to 1$ or $2 \to 2$ process. For a $2 \to 3$ or $2 \to 4$ $\W/\Z$ fusion process, it is set by the $\W/\Z$ mass scale, and for subprocesses 121 and 122 by the heavy-quark mass. \iteme{PARI(24) :} $Q^2$ of the outer hard-scattering subprocess; see {\tt PARI(23)}. \iteme{PARI(25) :} $Q$ scale used as maximum virtuality in parton showers. Is equal to {\tt PARI(23)}, except for deep-inelastic-scattering processes when {\tt MSTP(22)}$\geq 1$. \iteme{PARI(26) :} $Q^2$ scale in parton showers; see {\tt PARI(25)}. \iteme{PARI(31), PARI(32) :} the momentum fractions $x$ of the initial-state parton-shower initiators on side 1 and 2, respectively. \iteme{PARI(33), PARI(34) :} the momentum fractions $x$ taken by the partons at the hard interaction, as used e.g. in the parton-distribution functions. \iteme{PARI(35) :} Feynman-$x$, $x_F = x_1 - x_2 = ${\tt PARI(33)}$-${\tt PARI(34)}. \iteme{PARI(36) :} $\tau = \hat{s}/s = x_1 \, x_2 = ${\tt PARI(33)}$\times${\tt PARI(34)}. \iteme{PARI(37) :} $y = (1/2) \ln(x_1/x_2)$, i.e. rapidity of the hard-interaction subsystem in the c.m. frame of the event as a whole. \iteme{PARI(38) :} $\tau' = \hat{s}'/s = ${\tt PARI(20)/PARI(12)}. \iteme{PARI(39), PARI(40) :} the primordial $k_{\perp}$ values selected in the two beam remnants. \iteme{PARI(41) :} $\cos\hat{\theta}$, where $\hat{\theta}$ is the scattering angle of a $2 \to 2$ (or $2 \to 1 \to 2$) interaction, defined in the rest frame of the hard-scattering subsystem. \iteme{PARI(42) :} $x_{\perp}$, i.e. scaled transverse momentum of the hard-scattering subprocess, $x_{\perp} = 2 \hat{p}_{\perp}/E_{\mrm{cm}}$. \iteme{PARI(43), PARI(44) :} $x_{L3}$ and $x_{L4}$, i.e. longitudinal momentum fractions of the two scattered partons, in the range $-1 < x_{\mrm{L}} < 1$, in the c.m. frame of the event as a whole. \iteme{PARI(45), PARI(46) :} $x_3$ and $x_4$, i.e. scaled energy fractions of the two scattered partons, in the c.m. frame of the event as a whole. \iteme{PARI(47), PARI(48) :} $y^*_3$ and $y^*_4$, i.e. rapidities of the two scattered partons in the c.m. frame of the event as a whole. \iteme{PARI(49), PARI(50) :} $\eta^*_3$ and $\eta^*_4$, i.e. pseudorapidities of the two scattered partons in the c.m. frame of the event as a whole. \iteme{PARI(51), PARI(52) :} $\cos\theta^*_3$ and $\cos\theta^*_4$, i.e. cosines of the polar angles of the two scattered partons in the c.m. frame of the event as a whole. \iteme{PARI(53), PARI(54) :} $\theta^*_3$ and $\theta^*_4$, i.e. polar angles of the two scattered partons, defined in the range $0 < \theta^* < \pi$, in the c.m. frame of the event as a whole. \iteme{PARI(55), PARI(56) :} azimuthal angles $\phi^*_3$ and $\phi^*_4$ of the two scattered partons, defined in the range $-\pi < \phi^* < \pi$, in the c.m. frame of the event as a whole. \iteme{PARI(61) :} multiple interaction enhancement factor for current event. A large value corresponds to a central collision and a small value to a peripheral one. \iteme{PARI(65) :} sum of the transverse momenta of partons generated at the hardest interaction of the event, excluding initial- and final-state radiation, i.e. $2 \times${\tt PARI(17)}. \iteme{PARI(66) :} sum of the transverse momenta of all partons generated at the hardest interaction, including initial- and final-state radiation, resonance decay products, and primordial $k_{\perp}$. \iteme{PARI(67) :} sum of transverse momenta of partons generated at hard interactions, excluding the hardest one (see {\tt PARI(65)}), and also excluding initial- and final-state radiation. Is non-vanishing only in the multiple-interaction scenario. \iteme{PARI(68) :} sum of transverse momenta of all partons generated at hard interactions, excluding the hardest one (see {\tt PARI(66)}), but including initial- and final-state radiation. Is non-vanishing only in the multiple-interaction scenario. \iteme{PARI(69) :} sum of transverse momenta of all partons generated in hard interactions ({\tt PARI(66) + PARI(68)}) and, additionally, of all beam remnant partons. \iteme{PARI(71), PARI(72) :} sum of the momentum fractions $x$ taken by initial-state parton-shower initiators on side 1 and and side 2, excluding those of the hardest interaction. Is non-vanishing only in the multiple-interaction scenario. \iteme{PARI(73), PARI(74) :} sum of the momentum fractions $x$ taken by the partons at the hard interaction on side 1 and side 2, excluding those of the hardest interaction. Is non-vanishing only in the multiple-interaction scenario. \iteme{PARI(75), PARI(76) :} the $x$ value of a photon that branches into quarks or gluons, i.e. $x$ at interface between initial-state QED and QCD cascades. \iteme{PARI(77), PARI(78) :} the $\chi$ values selected for beam remnants that are split into two objects, describing how the energy is shared (see {\tt MSTP(92)} etc.); is vanishing if no splitting is needed. \iteme{PARI(81) :} size of the threshold factor (enhancement or suppression) in the latest event with heavy-flavour production; see {\tt MSTP(35)}. \iteme{PARI(91) :} average multiplicity $\br{n}$ of pile-up events, see {\tt MSTP(133)}. Only relevant for {\tt MSTP(133)=} 1 or 2. \iteme{PARI(92) :} average multiplicity $\langle n \rangle$ of pile-up events as actually simulated, i.e. with multiplicity = 0 events removed and the high-end tail truncated. Only relevant for {\tt MSTP(133)=} 1 or 2. \iteme{PARI(93) :} for {\tt MSTP(133)=1} it is the probability that a beam crossing will produce a pile-up event at all, i.e. that there will be at least one hadron--hadron interaction; for {\tt MSTP(133)=2} the probability that a beam crossing will produce a pile-up event with one hadron--hadron interaction of the desired rare type. \end{entry} \subsection{How to include external processes in PYTHIA} \label{ss:PYnewproc} Despite a large repertory of processes in {\Py}, the number of missing ones clearly is even larger, and with time this discrepancy is likely to increase. There are several reasons why it is not practicable to imagine a {\Py} which has `everything'. One is the amount of time it takes to implement a process for the single {\Py} author, compared with the rate of new cross section results produced by the rather larger matrix-element calculations community. Another is the length of currently produced matrix-element expressions, which would make the program very bulky. A third argument is that, whereas the phase space of $2 \to 1$ and $2 \to 2$ processes can be set up once and for all according to a reasonably flexible machinery, processes with more final-state particles are less easy to generate. To achieve a reasonable efficiency, it is necessary to tailor the phase-space selection procedure to the dynamics of the given process, and to the desired experimental cuts. If the desired subprocess is missing, it can be included into {\Py} as an `external' subprocess. In this section we will describe how it is possible to specify the partonic state of some hard-scattering process in an interface common block. {\Py} will read this common block, and add initial- and final-state showers, beam remnants and underlying events, fragmentation and decays, to build up an event in as much detail as an ordinary {\Py} one. You may also use {\Py} to mix events of different kinds, and to keep track of cross section statistics. You have to provide the matrix elements, the phase-space generator, and the storage of event information in the common block. First a minor comment, however. Some processes may be seen just as trivial modifications of already existing ones. For instance, you might want to add some extra term, corresponding to contact interactions, to the matrix elements of a {\Py} $2 \to 2$ process. In that case it is not necessary to go through the machinery below, but instead you can use the {\tt PYEVWT} routine to introduce an additional weight for the event, defined as the ratio of the modified to the unmodified differential cross sections. If you use the option {\tt MSTP(142)=2}, this weight is considered as part of the `true' cross section of the process, and the generation is changed accordingly. The more generic facility for including an external process is a bit more complicated, and involves two routines and one common block. All names contain {\tt UP}, which is short for User Process. If you want to include a new process, first you have to pick an unused subprocess number ISUB (see tables in section \ref{ss:ISUBcode}). For instance, the numbers 191--200 are currently unused, so this might be a logical place to put a new process. This number and the `title' of the process (plus {\tt SIGMAX}, to be described below) have to be given in to {\Py} in a subroutine call \drawbox{CALL PYUPIN(ISUB,TITLE,SIGMAX)}\label{p:PYUPIN} \boxsep \noindent before the call to {\tt PYINIT}. The {\tt TITLE} can be any character string up to 28 characters, e.g. \begin{verbatim} CALL PYUPIN(191,'g + g -> t + tbar + gamma',SIGMAX) \end{verbatim} The call to {\tt PYUPIN} tells the program that a process ISUB exists, but not that you want to generate it. This is done, as with normal processes, by setting {\tt MSUB(ISUB)=1} before the {\tt PYINIT} call. Once the event generation chain has been started and {\tt PYEVNT} is called to generate an event, this routine may in its turn call the routine {\tt PYUPEV}, which is the routine you must supply, in which the next event is selected. (A dummy copy of {\tt PYUPEV} has been included at the end of {\Py}; depending on the machine you may have to comment out this copy when you link your own.) The call arguments are \drawbox{CALL PYUPEV(ISUB,SIGEV)}\label{p:PYUPEV} \boxsep \noindent where {\tt ISUB} is given by {\tt PYEVNT}, while {\tt SIGEV} is to be calculated (see below) and returned to {\tt PYEVNT}. If there is only one user-defined process, then the {\tt ISUB} input is superfluous; otherwise it is necessary to branch to the relevant process. The {\tt SIGEV} variable is supposed to give the differential cross section ot the current event, times the phase-space volume within which events are generated, expressed in millibarns. This means that, in the limit that many events are generated, the average value of {\tt SIGEV} gives the total cross section of the simulated process. The {\tt SIGMAX} value, handed to {\Py} in the {\tt PYUPIN} call, is assumed to be the maximum value that {\tt SIGEV} will reach. Events will be accepted with a probability {\tt SIGEV/SIGMAX}, i.e. the acceptance/rejection of events according to differential cross section is done by {\tt PYEVNT}, not by the user. This means that the events that come out in the end all have unit weight, i.e. the user does not have to worry about events with different weights. It also allows several subprocesses to be generated together, in the proper mixture. Of course, the tricky part is that the differential cross section usually is strongly peaked in a few regions of the phase space, such that the average probability to accept an event, $\langle${\tt SIGEV}$\rangle /${\tt SIGMAX} is small. It may then be necessary to find a suitable set of transformed phase-space coordinates, for which the correspondingly transformed differential cross section is better behaved. To avoid unclarities, here is a more formal version of the two above paragraphs. Call $\d X$ the differential phase space, e.g. for a $2 \to 2$ process $\d X = \d x_1 \, \d x_2 \, \d \hat{t}$, where $x_1$ and $x_2$ are the momentum fractions carried by the two incoming partons and $\hat{t}$ the Mandelstam variable of the scattering. Call $\d \sigma / \d X$ the differential cross section of the process, e.g. for $2 \to 2$: $\d \sigma / \d X = \sum_{ij} f_i(x_1,Q^2) \, f_j(x_2,Q^2) \, \d \hat{\sigma}_{ij} / \d \hat{t}$, i.e. the product of parton distributions and hard-scattering matrix elements, summed over all allowed incoming flavours $i$ and $j$. The physical cross section that one then wants to generate is $\sigma = \int (\d \sigma / \d X) \, \d X$, where the integral is over the allowed phase-space volume. The event generation procedure consists of selecting an $X$ uniformly in $\d X$ and then evaluating the weight $\d \sigma / \d X$ at this point. {\tt SIGEV} is now simply {\tt SIGEV}$ = \d \sigma / \d X \, \int \d X$, i.e. the differential cross section times the considered volume of phase space. Clearly, when averaged over many events, {\tt SIGEV} will correctly estimate the desired cross section. If {\tt SIGEV} fluctuates too much, one may try to transform to new variables $X'$, where events are now picked accordingly to $\d X'$ and {\tt SIGEV}$ = \d \sigma / \d X' \, \int \d X'$. A warning. It is important that $X$ is indeed uniformly picked within the allowed phase space, alternatively that any Jacobians are properly taken into account. For instance, in the case above, one approach would be to pick $x_1$, $x_2$ and $\hat{t}$ uniformly in the ranges $0 < x_1 < 1$, $0 < x_2 < 1$, and $-s < \hat{t} < 0$, with full phase space volume $\int \d X = s$. The cross section would only be non-vanishing inside the physical region given by $-s x_1 x_2 < \hat{t}$ (in the massless case), i.e. Monte Carlo efficiency is likely to be low. However, if one were to choose $\hat{t}$ values only in the range $-\hat{s} < \hat{t} < 0$, small $\hat{s}$ values would be favoured, since the density of selected $\hat{t}$ values would be larger there. Without the use of a compensating Jacobian $\hat{s}/s$, an incorrect answer would be obtained. Alternatively, one could start out with a phase space like $\d X = \d x_1 \, \d x_2 \, \d (\cos\hat{\theta})$, where the limits decouple. Of course, the $\cos\hat{\theta}$ variable can be translated back into a $\hat{t}$, which will then always be in the desired range $-\hat{s} < \hat{t} < 0$. The transformation itself here gives the necessary Jacobian. If you do not know how big {\tt SIGMAX} is, you can put it to some very small value (but larger than zero, however) and do an exploratory run. When the program encounters events with {\tt SIGEV}$ > ${\tt SIGMAX}, a warning message is printed, which gives the new {\tt SIGMAX} that the program will use from then on. Hopefully such maximum violations only appear at the beginning of the run, and later {\tt SIGMAX} stabilizes to a level that can then be used as input for a second, correct run. If you want to do the event rejection yourself, simply put {\tt SIGEV} equal to {\tt SIGMAX}. In that case events will not be rejected by {\Py} (except if there is something else wrong with them). If {\tt SIGMAX} is the correct total cross section of the process, event mixing with other processes will still work fine. You could also decide not to reject any events, but to use weighted ones. In that case you can only have one ISUB switched on in a run, since the program will not know how to mix different kinds of events, and you cannot use {\Py} to do cross section statistics for you. Therefore you could, for instance, put {\tt SIGMAX = SIGEV = 1}, and use a common block to transfer event weight and other information from your {\tt PYUPEV} routine to your main program. In addition to the {\tt SIGEV} value returned for each event, it is also necessary to return the event itself. This is done via the common block \drawbox{COMMON/PYUPPR/NUP,KUP(20,7),PUP(20,5),NFUP,IFUP(10,2),% Q2UP(0:10)}\label{p:PYUPPR} \boxsep The first part closely parallels the standard event record in the {\tt LUJETS} common block, see section \ref{ss:evrec}, although with a few simplifications. The number {\tt NUP}\label{p:NUP} gives the number of particles involved in the process, where a particle may be a quark, a lepton, a gauge boson, or anything else. The first two are simply the two incoming particles that initiate the hard scattering, while the remaining {\tt NUP-2} are the outgoing particles from the hard process. For each particle {\tt I}, with $1 \leq ${\tt I}$ \leq ${\tt NUP}, the following information is stored: \begin{entry} \iteme{KUP(I,1) :}\label{p:KUP} is = 1 normally. However, if you put it = 2 that signifies intermediate states that are not to be treated by {\Py}, but are included only to make the event record easier to read. \iteme{KUP(I,2) :} is the flavour code of a particle, i.e. the two incoming partons for {\tt I} = 1 and 2, and the outgoing particles for {\tt I}$ \geq 3$. The flavour codes are the standard KF ones, as used elsewhere in the program. \iteme{KUP(I,3) :} may be used to indicate the position of a mother. Such information may again make the record more readable, but is not really needed, and so one may well put all {\tt KUP(I,3)=0}. \iteme{KUP(I,4) :} for a final-state parton which carries colour, {\tt KUP(I,4)} gives the position of the parton from which the colour comes; otherwise it must be 0. \iteme{KUP(I,5) :} for a final-state parton that carries anticolour, {\tt KUP(I,5)} gives the position of the parton from which the anticolour comes; else it must be 0. \iteme{KUP(I,6) :} for an initial-state parton that carries colour, {\tt KUP(I,6)} gives the position of the parton to which the colour goes; else it must be 0. \iteme{KUP(I,7) :} for an initial-state parton which carries anticolour, {\tt KUP(I,7)} gives the position of the parton to which the anticolour goes; else it must be 0. \iteme{PUP(I,1) :}\label{p:PUP} $p_x$, i.e. $x$ momentum. \iteme{PUP(I,2) :} $p_y$, i.e. $y$ momentum. \iteme{PUP(I,3) :} $p_z$, i.e. $z$ momentum. \iteme{PUP(I,4) :} $E$, i.e. energy. \iteme{PUP(I,5) :} $m$, i.e. mass. \end{entry} After this brief summary, we proceed with more details and examples. To illustrate the issue of documentation in {\tt KUP(I,1)} and {\tt KUP(I,3)}, consider the case of $\W^+$ production and decay to $\u \dbar$, maybe as part of a more complex process. The final-state particles clearly are $\u$ and $\dbar$, so the $\W^+$ need not be given at all, but it would make the event listing easier to read. Therefore one should add the $\W^+$, but with {\tt KUP(I,1)=2}. (If the $\W^+$ would have been added with {\tt KUP(I,1)=1}, it would later have been treated by {\Py}/{\Je}, which means it would have been allowed to decay once more.) If the $\W^+$ is in line 3, the $\u$ in 4 and the $\dbar$ in 5, one could further put {\tt KUP(4,3)=3} and {\tt KUP(5,3)=3} to indicate that the $\u$ and $\dbar$ in lines 4 and 5 come from the $\W^+$ stored in line 3. The switch {\tt MSTP(128)} works in the same way for user-defined processes as for ordinary ones, i.e. decay products of resonances can optionally be omitted from the documentation section of the event record, and history pointers can be set slightly differently. The information the program has at its disposal for this purpose is in {\tt KUP(I,3)}; an entry with this value non-zero is considered as a resonance decay product. The colour-flow information for coloured particles (quarks, gluons, leptoquarks, \ldots) is needed to set up parton showers and fragmentation properly. Sometimes many different colour flows are possible for one and the same process, as discussed in section \ref{sss:QCDjetclass}. It is up to you whether or not you will include all possible colour flows in the appropriate mixture, but at least you must pick some representative colour configuration. Consider e.g. the case of $\g(1) + \g(2) \to \q(3) + \qbar(4)$, where the numbers give the position in the array. It is clear the $\q$ must get its colour from either of the two gluons, which means there are (at least) two possibilities. Picking the $\q$ colour to come from gluon 1, one would thus write {\tt KUP(3,4)=1}, to be read `the colour of parton 3 comes from parton 1'. By implication therefore also {\tt KUP(1,6)=3}, i.e. `the colour of parton 1 goes to parton 3', i.e. the colour flow is bookkept doubly. The anticolour now must flow from parton 2 to parton 4, i.e. {\tt KUP(2,7)=4} and {\tt KUP(4,5)=2}. This completely specifies the colours of the $\q$ and $\qbar$, but not of the two gluons. In fact, one colour in the initial state `annihilates' between the $\g(1)$ and $\g(2)$, i.e. the anticolour of gluon 1 and the colour of gluon 2 match, which may be expressed by {\tt KUP(1,7)=2} and {\tt KUP(2,6)=1}. In other words colour/anticolour of an initial-state parton may either go to a final-state parton or to another initial-state parton. Correspondingly, the colour/anticolour of a final-state parton may come either from an initial-state parton or from another final-state parton. An example of the latter possibility is $\W$ decays, or generically the decay of any colour-singlet particle. (Thus a third colour flow above is represented by $\g \g \to \H^0 \to \q \qbar$, where no colour passes through the Higgs, and therefore colour flows between the two gluons and, separately, between the $\q$ and $\qbar$.) Storing of momenta should be straightforward, but a few comments must be made. Even if you ask, in the {\tt PYINIT} call, to have events generated in a fixed target or a user-specified frame, at intermediate stages {\Py} will still work in the c.m. frame of the two incoming beam particles, with the first beam moving in the $+z$ direction and the second in the $-z$ one. This c.m. frame must also be used when giving the momenta of the process. In addition, the two incoming partons in lines 1 and 2 are assumed massless. Therefore the initial-state partons are characterized only by the two energies {\tt P(1,4)} and {\tt P(2,4)}, with {\tt P(1,3) = P(1,4)}, {\tt P(2,3) = -P(2,4)}, and everything else is zero. In the final state, energies, momenta and masses are free, but must add up to give the same four-momentum as that of the initial state. All momenta are given in GeV, with the speed of light $c = 1$. The second part of the {\tt PYUPPR} common block is used to regulate the initial- and final-state showering, as follows: \begin{entry} \iteme{Q2UP(0) :}\label{p:Q2UP} $Q^2$ scale of initial-state showers. \iteme{NFUP :}\label{p:NFUP} number of parton pairs that undergo final-state showers. \iteme{IFUP(IF,1), IFUP(IF,2) :}\label{p:IFUP} positions of the two partons of a final-state showering pair, where the index {\tt IF} runs between 1 and {\tt NFUP}. \iteme{Q2UP(IF) :} the $Q^2$ scale of the final-state shower between parton pair {\tt IF} above. \end{entry} If you do not want any showering at all, you can put {\tt MSTP(61)=0} and {\tt MSTP(71)=0}, and then you do not have to give the above quantities. In general the scale choices $Q^2$ are not unique, which means that some guesswork is involved. Since the showers add extra partonic activity at mass scales below the mentioned $Q^2$ choices, the {\tt Q2UP} should be of the order of the phase-space cut-offs, so as to provide a reasonably smooth joining between partonic activity from matrix elements and that from showers. There are a few cases where choices are rather easy. In the decay of any $s$-channel colour neutral state, such as a $\W^{\pm}$, the $Q^2$ scale of final-state showers is just set by the squared mass of the resonance. For initial-state radiation, {\tt Q2UP(0)} should be about the same as the $Q^2$ scale used for the evaluation of parton distributions for the hard process, up to some factor of order unity. (One frequent choice for this factor would be 4, if your parton-distribution scale is something like the squared transverse momentum, simply because $m^2$ is of order $4 \pT^2$.) The `parton'-shower evolution actually also can include photon emission off quarks and leptons, if the shower switches are properly set. It is not possible to define only one particle in the above arrays, since it would then not be possible to conserve energy and momentum in the shower. You can very well have a pair where only one of the two can branch, however. For instance, in a $\g \gamma$ final state, only the gluon can shower, but the photon can lose energy to the gluon in such a way that the gluon branchings becomes possible. Currently, it is not possible to do showering where three or more final-state particles are involved at the same time. This may be added at a later stage. It is therefore necessary to subdivide suitably into pairs, and maybe leave some (especially colour-neutral) particles unshowered. You are free to make use of whatever tools you want in your {\tt PYUPEV} routine, and normally there would be no contact with the rest of {\Py}, except as described above. However, you may want to use some of the tools already available. One attractive possibility is to use {\tt PYSTFU} for parton-distribution-function evaluation. Other possible tools could be {\tt RLU} for random-number generation, {\tt ULALPS} for $\alphas$ evaluation, {\tt ULALEM} for evaluation of a running $\alphaem$, and maybe a few more. We end with a few comments on anomalous situations. In some cases one may want to decide, inside {\tt PYUPEV}, when to stop the event-generation loop. This is the case, for instance, if event configurations are read in from a file, and the end of the file is reached. One might be tempted just to put {\tt SIGEV=0} when this happens. Then {\tt PYEVNT} will discard the event, as part of the matrix-element-weighting procedure. However, next {\tt PYEVNT} will generate another event, which normally means a new request to {\tt PYUPEV}, so one does not really get out of the loop. Instead you should put {\tt NUP=0}. If the program encounters this value at a return from {\tt PYUPEV}, then it will also exit from {\tt PYEVNT}, without incrementing the counters for the number of events generated. It is then up to you to have a check on this condition in your main event-generation loop. This you do either by looking at {\tt NUP} or at {\tt MSTI(51)}; the latter is set to 1 if no event was generated. It may also happen that a user-defined configuration fails elsewhere in the {\tt PYEVNT} call. For instance, the beam-remnant treatment occasionally encounters situations it cannot handle, wherefore the hard interaction is rejected and a new one generated. This happens also with ordinary (not user-defined) events, and usually comes about as a consequence of the initial-state radiation description leaving too little energy for the remnant. If the same hard scattering were to be used as input for a new initial-state radiation and beam-remnant attempt, it could then work fine. There is a possibility to give events that chance, as follows. {\tt MSTI(52)} counts the number of times a hard-scattering configuration has failed to date. If you come in to {\tt PYUPEV} with {\tt MSTI(52)} non-vanishing, this means that the latest configuration failed. So long as the contents of the {\tt PYUPPR} common block are not changed, such an event may be given another try. For instance, a line \begin{verbatim} IF(MSTI(52).GE.1.AND.MSTI(52).LE.4) RETURN \end{verbatim} at the beginning of {\tt PYUPEV} will give each event up to five tries; thereafter a new one would be generated as usual. Note that the counter for the number of events is updated at each new try. The fraction of failed configurations is given in the bottom line of the {\tt PYSTAT(1)} table. The above comment only refers to very rare occurrences (less than one in a hundred), which are not errors in a strict sense; for instance, they do not produce any error messages on output. If you get warnings and error messages that the program does not understand the flavour codes or cannot reconstruct the colour flows, it is due to faults of yours, and giving such events more tries is not going to help. \subsection{How to run PYTHIA with varying energies} \label{ss:PYvaren} It is possible to use {\Py} in a mode where the energy can be varied from one event to the next, without the need to reinitialize with a new {\tt PYINIT} call. This allows a significant speed-up of execution, although it is not as fast as running at a fixed energy. It can not be used for everything --- we will come to the fine print at the end --- but it should be applicable for most tasks. The master switch to access this possibility is in {\tt MSTP(171)}. By default it is off, so you must set {\tt MSTP(171)=1} before initialization. There are two submodes of running, with {\tt MSTP(172)} being 1 or 2. In the former mode, {\Py} will generate an event at the requested energy. This means that you have to know which energy you want beforehand. In the latter mode, {\Py} will often return without having generated an event --- with flag {\tt MSTI(61)=1} to signal that --- and you are then requested to give a new energy. The energy spectrum of accepted events will then, in the end, be your naive input spectrum weighted with the cross-section of the processes you study. We will come back to this. The energy can be varied, whichever frame is given in the {\tt PYINIT} call. When the frame is {\tt 'CMS'}, {\tt PARP(171)} should be filled with the fractional energy of each event, i.e. $E_{\mrm{cm}} =${\tt PARP(171)}$\times${\tt WIN}, where {\tt WIN} is the nominal c.m. energy of the {\tt PYINIT} call. Here {\tt PARP(171)} should normally be smaller than unity, i.e. initialization should be done at the maximum energy to be encountered. For the {\tt 'FIXT'} frame, {\tt PARP(171)} should be filled by the fractional beam energy of that one, i.e. $E_{beam} = ${\tt PARP(171)}$\times${\tt WIN}. For the {\tt 'USER'}, {\tt 'FOUR'} and {\tt 'FIVE'} options, the two four-momenta are given in for each event in the same format as used for the {\tt PYINIT} call. To illustrate the use of the {\tt MSTP(172)=2} facility, consider the case of beamstrahlung in $\ee$ linear colliders. This is just for convenience; what is said here can be translated easily into other situations. Assume that the beam spectrum is given by $D(z)$, where $z$ is the fraction retained by the original $\e$ after beamstrahlung. Therefore $0 \leq z \leq 1$ and the integral of $D(z)$ is unity. This is not perfectly general; one could imagine branchings $\e^- \to \e^- \gamma \to \e^-\e^+\e^-$, which gives a multiplication in the number of beam particles. This could either be expressed in terms of a $D(z)$ with integral larger than unity or in terms of an increased luminosity. We will assume the latter, and use $D(z)$ properly normalized. Given a nominal $s = 4E_{beam}^2$, the actual $s'$ after beamstrahlung is given by $s' = z_1 z_2 s$. For a process with a cross section $\sigma(s)$ the total cross section is then \begin{equation} \sigma_{tot} = \int_0^1 \int_0^1 D(z_1) \, D(z_2) \sigma(z_1 z_2 s) \, \d z_1 \, \d z_2~. \end{equation} The cross section $\sigma$ may in itself be an integral over a number of additional phase space variables. If the maximum of the differential cross section is known, a correct procedure to generate events is \begin{Enumerate} \item pick $z_1$ and $z_2$ according to $D(z_1) \, \d z_1$ and $D(z_2) \, \d z_2$, respectively; \item pick a set of phase space variables of the process, for the given $s'$ of the event; \item evaluate $\sigma(s')$ and compare with $\sigma_{max}$; \item if event is rejected, then return to step 1 to generate new variables; \item else continue the generation to give a complete event. \end{Enumerate} You as a user are assumed to take care of step 1, and present the resulting kinematics with incoming $\e^+$ and $\e^-$ of varying energy. Thereafter {\Py} will do steps 2--5, and either return an event or put {\tt MSTI(61)=1} to signal failure in step 4. The maximization procedure does search in phase space to find $\sigma_{max}$, but it does not vary the $s'$ energy in this process. Therefore the maximum search in the {\tt PYINIT} call should be performed where the cross section is largest. For processes with increasing cross section as a function of energy this means at the largest energy that will ever be encountered, i.e. $s' = s$ in the case above. This is the `standard' case, but often one encounters other behaviours, where more complicated procedures are needed. One such case would be the process $\ee \to \Z^{*0} \to \Z^0 \H^0$, which is known to have a cross section that increases near the threshold but is decreasing asymptotically. If one already knows that the maximum, for a given Higgs mass, appears at 300 GeV, say, then the {\tt PYINIT} call should be made with that energy, even if subsequently one will be generating events for a 500 GeV collider. In general, it may be necessary to modify the selection of $z_1$ and $z_2$ and assign a compensating event weight. For instance, consider a process with a cross section behaving roughly like $1/s$. Then the $\sigma_{tot}$ expression above may be rewritten as \begin{equation} \sigma_{tot} = \int_0^1 \int_0^1 \frac{D(z_1)}{z_1} \, \frac{D(z_2)}{z_2} \, z_1 z_2 \sigma(z_1 z_2 s) \, \d z_1 \, \d z_2 ~. \end{equation} The expression $z_1 z_2 \sigma(s')$ is now essentially flat in $s'$, i.e. not only can $\sigma_{max}$ be found at a convenient energy such as the maximum one, but additionally the {\Py} generation efficiency (the likelihood of surviving step 4) is greatly enhanced. The price to be paid is that $z$ has to be selected according to $D(z)/z$ rather than according to $D(z)$. Note that $D(z)/z$ is not normalized to unity. One therefore needs to define \begin{equation} {\cal I}_D = \int_0^1 \frac{D(z)}{z} \, \d z ~, \end{equation} and a properly normalized \begin{equation} D'(z) = \frac{1}{I_D} \, \frac{D(z)}{z} ~. \end{equation} Then \begin{equation} \sigma_{tot} = \int_0^1 \int_0^1 D'(z_1) \, D'(z_2) \, {\cal I}_D^2 \, z_1 z_2 \sigma(z_1 z_2 s) \, \d z_1 \, \d z_2 ~. \end{equation} Therefore the proper event weight is ${\cal I}_D^2 \, z_1 z_2$. This weight should be stored, for each event, in {\tt PARP(173)}. Additionally you must put {\tt MSTP(173)=1} for the program to make use of weights at all. Often $D(z)$ are not known analytically; therefore ${\cal I}_D$ is also not known beforehand, but may have to be evaluated (by you) during the course of the run. Then you should just use the weight $z_1 z_2$ in {\tt PARP(173)} and do the overall normalization yourself in the end. Only the cross sections are affected by this, the events themselves still are properly distributed in $s'$ and internal phase space. Above it has been assumed tacitly that $D(z) \to 0$ for $z \to 0$. If not, $D(z)/z$ is divergent, and it is not possible to define a properly normalized $D'(z) = D(z)/z$. If the cross section is truly diverging like $1/s$, then a $D(z)$ which is nonvanishing for $z \to 0$ does imply an infinite total cross section, whichever way things are considered. In cases like that, it is necessary to impose a lower cut on $z$, based on some physics or detector consideration. Some such cut is anyway needed to keep away from the minimum c.m. energy required for {\Py} events, see {\tt PARP(2)}. The most difficult cases are those with a very narrow and high peak, such as the $\Z^0$. One could initialize at the energy of maximum cross section and use $D(z)$ as is, but efficiency might turn out to be very low. One might then be tempted to do more complicated transforms of the kind illustrated above. As a rule it is then convenient to work in the variables $\tau_z = z_1 z_2$ and $y_z = (1/2) \ln (z_1/z_2)$, cf. subsection \ref{ss:kinemtwo}. Clearly, the better the behaviour of the cross section can be modelled in the choice of $z_1$ and $z_2$, the better the overall event generation efficieny. Even under the best of circumstances, the efficiency will still be lower than for runs with fix energy. There is also a non-negligible time overhead for using variable energies in the first place, from kinematics reconstruction and (in part) from the phase space selection. One should therefore not use variable energies when not needed, and not use a large range of energies $\sqrt{s'}$ if in the end only a smaller range is of experimental interest. This facility may be combined with most other aspects of the program. For instance, it is possible to simulate beamstrahlung as above and still include bremsstrahlung with {\tt MSTP(11)=1}. Further, one may multiply the overall event weight of {\tt PARP(173)} with a kinematics-dependent weight given by {\tt PYEVWT}, although it is not recommended (since the chances of making a mistake are also multiplied). However, a few things do {\em not} work. \begin{Itemize} \item It is not possible to use pile-up events, i.e. you must have {\tt MSTP(131)=0}. \item The possibility of giving in your own cross-section optimization coefficients, option {\tt MSTP(121)=2}, would require more input than with fixed energies, and this option should therefore not be used. You can still use {\tt MSTP(121)=1}, however. \item The multiple interactions scenario with {\tt MSTP(82)}$\geq 2$ only works approximately for energies different from the initialization one. If the c.m. energy spread is smaller than a factor 2, say, the approximation should be reasonable, but if the spread is larger one may have to subdivide into subruns of different energy bins. The initialization should be made at the largest energy to be encountered --- whenever multiple interactions are possible (i.e. for incoming hadrons and resolved photons) this is where the cross sections are largest anyway, and so this is no further constraint. There is no simple possibility to change {\tt PARP(82)} during the course of the run, i.e. an energy-independent $p_{\perp 0}$ must be assumed. The default option {\tt MSTP(82)=1} works fine, i.e. does not suffer from the constraints above. If so desired, $\pTmin=${\tt PARP(81)} can be set differently for each event, as a function of c.m. energy. Initialization should then be done with {\tt PARP(81)} as low as it is ever supposed to become. \end{Itemize} \subsection{Other Routines and Common Blocks} The subroutines and common blocks that you will come in direct contact with have already been described. A number of other routines and common blocks exist, and those not described elsewhere are here briefly listed for the sake of completeness. \begin{entry} \iteme{SUBROUTINE PYINKI(CHFRAM,CHBEAM,CHTARG,WIN) :}\label{p:PYINKI} to initialize the kinematics given by the two incoming particles. \iteme{SUBROUTINE PYINRE :}\label{p:PYINRE} to initialize the widths and effective widths of resonances. \iteme{SUBROUTINE PYXTOT :}\label{p:PYXTOT} to give the parametrized total, double diffractive, single diffractive and elastic cross sections for different energies and colliding hadrons. \iteme{SUBROUTINE PYMAXI :}\label{p:PYMAXI} to find optimal coefficients {\tt COEF} for the selection of kinematical variables, and to find the related maxima for the differential cross section times Jacobian factors, for each of the subprocesses included. \iteme{SUBROUTINE PYEVKI :}\label{p:PYEVKI} to find event kinematics for each individual event, when variable energies are used. \iteme{SUBROUTINE PYPILE(MPILE) :}\label{p:PYPILE} to determine the number of pile-up events, i.e. events appearing in the same beam--beam crossing. \iteme{SUBROUTINE PYRAND :}\label{p:PYRAND} to generate the quantities characterizing a hard scattering on the parton level, according to the relevant matrix elements. \iteme{SUBROUTINE PYSCAT :}\label{p:PYSCAT} to find outgoing flavours and to set up the kinematics and colour flow of the hard scattering. \iteme{SUBROUTINE PYRESD :}\label{p:PYRESD} to allow resonances to decay, including chains of successive decays and parton showers. \iteme{SUBROUTINE PYMULT(MMUL) :}\label{p:PYMULT} to generate semi-hard interactions according to the multiple interaction formalism. \iteme{SUBROUTINE PYREMN(IPU1,IPU2) :}\label{p:PYREMN} to add on target remnants and include primordial $k_{\perp}$. \iteme{SUBROUTINE PYDIFF :}\label{p:PYDIFF} to handle diffractive and elastic scattering events. \iteme{SUBROUTINE PYDOCU :}\label{p:PYDOCU} to compute cross sections of processes, based on current Monte Carlo statistics, and to store event information in the {\tt MSTI} and {\tt PARI} arrays. \iteme{SUBROUTINE PYWIDT(KFLR,SH,WDTP,WDTE) :}\label{p:PYWIDT} to calculate widths and effective widths of resonances. \iteme{SUBROUTINE PYOFSH(MOFSH,KFMO,KFD1,KFD2,PMMO,RET1,RET2) :}% \label{p:PYOFSH} to calculate partial widths into channels off the mass shell, and to select correlated masses of resonance pairs. \iteme{SUBROUTINE PYKLIM(ILIM) :}\label{p:PYKLIM} to calculate allowed kinematical limits. \iteme{SUBROUTINE PYKMAP(IVAR,MVAR,VVAR) :}\label{p:PYKMAP} to calculate the value of a kinematical variable when this is selected according to one of the simple pieces. \iteme{SUBROUTINE PYSIGH(NCHN,SIGS) :}\label{p:PYSIGH} to give the differential cross section (multiplied by the relevant Jacobians) for a given subprocess and kinematical setup. \iteme{SUBROUTINE PYSTFL(KF,X,Q2,XPQ) :}\label{p:PYSTFL} to give parton distributions for $\p$ and $\n$ in the option with modified behaviour at small $Q^2$ and $x$, see {\tt MSTP(57)}. \iteme{SUBROUTINE PYSTFU(KF,X,Q2,XPQ) :}\label{p:PYSTFU} to give parton-distribution functions (multiplied by $x$, i.e. $x f_i(x,Q^2)$) for an arbitrary particle (of those recognized by {\Py}). Generic driver routine for the following, specialized ones. \begin{subentry} \iteme{KF :} flavour of probed particle, according to KF code. \iteme{X :} $x$ value at which to evaluate parton distributions. \iteme{Q2 :} $Q^2$ scale at which to evaluate parton distributions. \iteme{XPQ :} array of dimensions {\tt XPQ(-25:25)}, which contains the evaluated parton distributions $x f_i(x,Q^2)$. Components $i$ ordered according to standard KF code; additionally the gluon is found in position 0 as well as 21 (for historical reasons). \end{subentry} \iteme{SUBROUTINE PYSTEL(X,Q2,XPEL) :}\label{p:PYSTEL} to give electron parton distributions. \iteme{SUBROUTINE PYSTGA(X,Q2,XPGA) :}\label{p:PYSTGA} to give photon parton distributions. \iteme{SUBROUTINE PYSTAG(X,Q2,XPGA) :}\label{p:PYSTAG} interface to give anomalous part of photon parton distributions. \iteme{SUBROUTINE PYSTGS(X,Q2,P2,ALAM,NF,PDFUNC) :}\label{p:PYSTGS} to give anomalous part of photon parton distribution; double precision routine written by Gerhard Schuler. \iteme{FUNCTION PYDILN(X) :}\label{p:PYDILN} to evaluate dilogarithm function in double precision. \iteme{SUBROUTINE PYSTPI(X,Q2,XPPI) :}\label{p:PYSTPI} to give pion parton distributions. \iteme{SUBROUTINE PYSTPR(X,Q2,XPPR) :}\label{p:PYSTPR} to give proton parton distributions. \iteme{FUNCTION PYHFTH(SH,SQM,FRATT) :}\label{p:PYHFTH} to give heavy-flavour threshold factor in matrix elements. \iteme{SUBROUTINE PYSPLI(KF,KFLIN,KFLCH,KFLSP) :}\label{p:PYSPLI} to give hadron remnant or remnants left behind when the reacting parton is kicked out. \iteme{FUNCTION PYGAMM(X) :}\label{p:PYGAMM} to give the value of the ordinary $\Gamma (x)$ function (used in some parton-distribution parametrizations). \iteme{SUBROUTINE PYWAUX(IAUX,EPS,WRE,WIM) :}\label{p:PYWAUX} to evaluate the two auxiliary functions $W_1$ and $W_2$ appearing in the cross section expressions in {\tt PYSIGH}. \iteme{SUBROUTINE PYI3AU(EPS,RAT,Y3RE,Y3IM) :}\label{p:PYI3AU} to evaluate the auxiliary function $I_3$ appearing in the cross section expressions in {\tt PYSIGH}. \iteme{FUNCTION PYSPEN(XREIN,XIMIN,IREIM) :}\label{p:PYSPEN} to calculate the real and imaginary part of the Spence function. \iteme{SUBROUTINE PYQQBH(WTQQBH) :}\label{p:PYQQBH} to calculate matrix elements for the two processes $\g \g \to \Q \Qbar \H^0$ and $\q \qbar \to \Q \Qbar \H^0$. \iteme{BLOCK DATA PYDATA :}\label{p:PYDATA} to give sensible default values to all status codes and parameters. \end{entry} \drawbox{COMMON/PYINT1/MINT(400),VINT(400)}\label{p:PYINT1} \begin{entry} \itemc{Purpose:} to collect a host of integer- and real-valued variables used internally in the program during the initialization and/or event generation stage. These variables must not be changed by you. \iteme{MINT(1) :}\label{p:MINT} specifies the general type of subprocess that has occurred, according to the ISUB code given in section \ref{ss:ISUBcode}. \iteme{MINT(2) :} whenever {\tt MINT(1)} (together with {\tt MINT(15)} and {\tt MINT(16)}) are not sufficient to specify the type of process uniquely, {\tt MINT(2)} provides an ordering of the different possibilities, see {\tt MSTI(2)}. \iteme{MINT(3) :} number of partons produced in the hard interactions, i.e. the number $n$ of the $2 \to n$ matrix elements used; is sometimes 3 or 4 when a basic $2 \to 1$ or $2 \to 2$ process has been folded with two $1 \to 2$ initial branchings (like $\q \q' \to \q'' \q''' \H^0$). \iteme{MINT(4) :} number of documentation lines at the beginning of the common block {\tt LUJETS} that are given with {\tt K(I,1)=21}; 0 for {\tt MSTP(125)=0}. \iteme{MINT(5) :} number of events generated to date in current run. \iteme{MINT(6) :} current frame of event (see {\tt MSTP(124)} for possible values). \iteme{MINT(7), MINT(8) :} line number for documentation of outgoing partons/particles from hard scattering for $2 \to 2$ or $2 \to 1 \to 2$ processes (else = 0). \iteme{MINT(10) :} is 1 if cross section maximum was violated in current event, and 0 if not. \iteme{MINT(11) :} KF flavour code for beam (side 1) particle. \iteme{MINT(12) :} KF flavour code for target (side 2) particle. \iteme{MINT(13), MINT(14) :} KF flavour codes for side 1 and side 2 initial-state shower initiators. \iteme{MINT(15), MINT(16) :} KF flavour codes for side 1 and side 2 incoming partons to the hard interaction. \iteme{MINT(17), MINT(18) :} flag to signal if particle on side 1 or side 2 has been scattered diffractively; 0 if no, 1 if yes. \iteme{MINT(19), MINT(20) :} flag to signal initial-state structure with parton inside photon inside electron on side 1 or side 2; 0 if no, 1 if yes. \iteme{MINT(21) - MINT(24) :} KF flavour codes for outgoing partons from the hard interaction. The number of positions actually used is process-dependent, see {\tt MINT(3)}; trailing positions not used are set = 0. \iteme{MINT(25), MINT(26) :} KF flavour codes of the products in the decay of a single $s$-channel resonance formed in the hard interaction. Are thus only used when {\tt MINT(3)=1} and the resonance is allowed to decay. \iteme{MINT(31) :} number of hard or semi-hard scatterings that occurred in the current event in the multiple-interaction scenario; is = 0 for a low-$\pT$ event. \iteme{MINT(35) :} in a true $2 \to 3$ process, where one particle is a resonance with decay channel selected already before the {\tt PYRESD} call, the decay channel number (in the {\tt /LUDAT3/} numbering) is stored here. \iteme{MINT(41), MINT(42) :} type of incoming beam or target particle; 1 for lepton and 2 for hadron. A photon counts as a lepton if it is not resolved ({\tt MSTP(14)=0}) and as a hadron if it is resolved ({\tt MSTP(14)}$\geq 1$). \iteme{MINT(43) :} combination of incoming beam and target particles. A photon counts as a hadron. \begin{subentry} \iteme{= 1 :} lepton on lepton. \iteme{= 2 :} lepton on hadron. \iteme{= 3 :} hadron on lepton. \iteme{= 4 :} hadron on hadron. \end{subentry} \iteme{MINT(44) :} as {\tt MINT(43)}, but a photon counts as a lepton. \iteme{MINT(45), MINT(46) :} structure of incoming beam and target particles. \begin{subentry} \iteme{= 1 :} no internal structure, i.e. an electron or photon carrying the full beam energy. \iteme{= 2 :} defined with parton distributions that are not peaked at $x = 1$, i.e. a hadron or a resolved photon. \iteme{= 3 :} defined with parton distributions that are peaked at $x = 1$, i.e. a resolved electron. \end{subentry} \iteme{MINT(47) :} combination of incoming beam- and target-particle parton-distribution function types. \begin{subentry} \iteme{= 1 :} no parton distribution either for beam or target. \iteme{= 2 :} parton distributions for target but not for beam. \iteme{= 3 :} parton distributions for beam but not for target. \iteme{= 4 :} parton distributions for both beam and target, but not both peaked at $x = 1$. \iteme{= 5 :} parton distributions for both beam and target, with both peaked at $x = 1$. \end{subentry} \iteme{MINT(48) :} total number of subprocesses switched on. \iteme{MINT(49) :} number of subprocesses that are switched on, apart from elastic scattering and single, double and central diffractive. \iteme{MINT(50) :} combination of incoming particles from a multiple interactions point of view. \begin{subentry} \iteme{= 0 :} the total cross section is not known; therefore no multiple interactions are possible. \iteme{= 1 :} the total cross section is known; therefore multiple interactions are possible if switched on. \end{subentry} \iteme{MINT(51) :} internal flag that event failed cuts. \begin{subentry} \iteme{= 0 :} no problem. \iteme{= 1 :} event failed; new one to be generated. \iteme{= 2 :} event failed; no new event is to be generated but instead control is to be given back to used. Is intended for user-defined processes, when {\tt NUP=0}. \end{subentry} \iteme{MINT(52) :} internal counter for number of lines used (in {\tt /LUJETS/}) before multiple interactions are considered. \iteme{MINT(53) :} internal counter for number of lines used (in {\tt /LUJETS/}) before beam remnants are considered. \iteme{MINT(55) :} the heaviest new flavour switched on for QCD processes, specifically the flavour to be generated for ISUB = 81, 82, 83 or 84. \iteme{MINT(56) :} the heaviest new flavour switched on for QED processes, specifically for ISUB = 85. Note that, unlike {\tt MINT(55)}, the heaviest flavour may here be a lepton, and that heavy means the one with largest KF code. \iteme{MINT(57) :} number of times the beam remnant treatment has failed, and the same basic kinematical setup is used to produce a new parton shower evolution and beam remnant set. Mainly used in leptoproduction, for the option when $x$ and $Q^2$ are to be preserved. \iteme{MINT(61) :} internal switch for the mode of operation of resonance width calculations in {\tt PYWIDT} for $\gammaZ$ or $\gamma^*/\Z^0/\Z'^0$. \begin{subentry} \iteme{= 0 :} without reference to initial-state flavours. \iteme{= 1 :} with reference to given initial-state flavours. \iteme{= 2 :} for given final-state flavours. \end{subentry} \iteme{MINT(62) :} internal switch for use at initialization of $\H^0$ width. \begin{subentry} \iteme{= 0 :} use widths into $\Z \Z^*$ or $\W \W^*$ calculated before. \iteme{= 1 :} evaluate widths into $\Z \Z^*$ or $\W \W^*$ for current Higgs mass. \end{subentry} \iteme{MINT(65) :} internal switch to indicate initialization without specified reaction. \begin{subentry} \iteme{= 0 :} normal initialization. \iteme{= 1 :} initialization with argument {\tt 'none'} in {\tt PYINIT} call. \end{subentry} \iteme{MINT(71) :} switch to tell whether current process is singular for $\pT \to 0$ or not. \begin{subentry} \iteme{= 0 :} non-singular process, i.e. proceeding via an $s$-channel resonance or with both products having a mass above {\tt CKIN(6)}. \iteme{= 1 :} singular process. \end{subentry} \iteme{MINT(72) :} number of $s$-channel resonances that may contribute to the cross section. \iteme{MINT(73) :} KF code of first $s$-channel resonance; 0 if there is none. \iteme{MINT(74) :} KF code of second $s$-channel resonance; 0 if there is none. \iteme{MINT(81) :} number of selected pile-up events. \iteme{MINT(82) :} sequence number of currently considered pile-up event. \iteme{MINT(83) :} number of lines in the event record already filled by previously considered pile-up events. \iteme{MINT(84) :} {\tt MINT(83) + MSTP(126)}, i.e. number of lines already filled by previously considered events plus number of lines to be kept free for event documentation. \iteme{MINT(91) :} is 1 for a lepton--hadron event and 0 else. Used to determine whether a {\tt PYFRAM(3)} call is possible. \iteme{MINT(92) :} is used to denote region in $(x,Q^2)$ plane when {\tt MSTP(57)=2}, according to numbering in \cite{Sch93a}. Simply put, 0 means that the modified proton parton distributions were not used, 1 large $x$ and $Q^2$, 2 small $Q^2$ but large $x$, 3 small $x$ but large $Q^2$ and 4 small $x$ and $Q^2$. \iteme{MINT(101), MINT(102) :} is normally 1, but is 4 when a resolved photon (appearing on side 1 or 2) can be represented by either of the four vector mesons $\rho^0$, $\omega$, $\phi$ and $\Jpsi$. \iteme{MINT(103), MINT(104) :} KF flavour code for the two incoming particles, i.e. the same as {\tt MINT(11)} and {\tt MINT(12)}. The exception is when a resolved photon is represented by a vector meson (a $\rho^0$, $\omega$ or $\phi$). Then the code of the vector meson is given. \iteme{MINT(107) :} normally 0, but is set 1 for the anomalous class, {\tt MSTP(14)=3}, of $\gamma\p$ events. It signals the need to select a virtuality scale for the branching $\gamma \to \q\qbar$. \iteme{MINT(111) :} the frame given in {\tt PYINIT} call, 0--5 for {\tt 'NONE'}, {\tt 'CMS'}, {\tt 'FIXT'}, {\tt 'USER'}, {\tt 'FOUR'} and {\tt 'FIVE'}, respectively. \boxsep \iteme{VINT(1) :}\label{p:VINT} $E_{\mrm{cm}}$, c.m. energy. \iteme{VINT(2) :} $s$ ($=E_{\mrm{cm}}^2$) squared mass of complete system. \iteme{VINT(3) :} mass of beam particle. \iteme{VINT(4) :} mass of target particle. \iteme{VINT(5) :} momentum of beam (and target) particle in c.m. frame. \iteme{VINT(6) - VINT(10) :} $\theta$, $\varphi$ and \mbox{{\boldmath $\beta$}} for rotation and boost from c.m. frame to user-specified frame. \iteme{VINT(11) :} $\tau_{\mmin}$. \iteme{VINT(12) :} $y_{\mmin}$. \iteme{VINT(13) :} $\cos\hat{\theta}_{\mmin}$ for $\cos\hat{\theta} \leq 0$. \iteme{VINT(14) :} $\cos\hat{\theta}_{\mmin}$ for $\cos\hat{\theta} \geq 0$. \iteme{VINT(15) :} $x^2_{\perp \mmin}$. \iteme{VINT(16) :} $\tau'_{\mmin}$. \iteme{VINT(21) :} $\tau$. \iteme{VINT(22) :} $y$. \iteme{VINT(23) :} $\cos\hat{\theta}$. \iteme{VINT(24) :} $\varphi$ (azimuthal angle). \iteme{VINT(25) :} $x_{\perp}^2$. \iteme{VINT(26) :} $\tau'$. \iteme{VINT(31) :} $\tau_{\mmax}$. \iteme{VINT(32) :} $y_{\mmax}$. \iteme{VINT(33) :} $\cos\hat{\theta}_{\mmax}$ for $\cos\hat{\theta} \leq 0$. \iteme{VINT(34) :} $\cos\hat{\theta}_{\mmax}$ for $\cos\hat{\theta} \geq 0$. \iteme{VINT(35) :} $x^2_{\perp \mmax}$. \iteme{VINT(36) :} $\tau'_{\mmax}$. \iteme{VINT(41), VINT(42) :} the momentum fractions $x$ taken by the partons at the hard interaction, as used e.g. in the parton-distribution functions. \iteme{VINT(43) :} $\hat{m} = \sqrt{\hat{s}}$, mass of hard-scattering subsystem. \iteme{VINT(44) :} $\hat{s}$ of the hard subprocess ($2 \to 2$ or $2 \to 1$). \iteme{VINT(45) :} $\hat{t}$ of the hard subprocess ($2 \to 2$ or $2 \to 1 \to 2$). \iteme{VINT(46) :} $\hat{u}$ of the hard subprocess ($2 \to 2$ or $2 \to 1 \to 2$). \iteme{VINT(47) :} $\hat{p}_{\perp}$ of the hard subprocess ($2 \to 2$ or $2 \to 1 \to 2$), i.e. transverse momentum evaluated in the rest frame of the scattering. \iteme{VINT(48) :} $\hat{p}_{\perp}^2$ of the hard subprocess; see {\tt VINT(47)}. \iteme{VINT(49) :} $\hat{m}'$, the mass of the complete three- or four-body final state in $2 \to 3$ or $2 \to 4$ processes. \iteme{VINT(50) :} $\hat{s}' = \hat{m}'^2$; see {\tt VINT(49)}. \iteme{VINT(51) :} $Q$ of the hard subprocess. The exact definition is process-dependent, see {\tt MSTP(32)}. \iteme{VINT(52) :} $Q^2$ of the hard subprocess; see {\tt VINT(51)}. \iteme{VINT(53) :} $Q$ of the outer hard-scattering subprocess. Agrees with {\tt VINT(51)} for a $2 \to 1$ or $2 \to 2$ process. For a $2 \to 3$ or $2 \to 4$ $\W/\Z$ fusion process, it is set by the $\W/\Z$ mass scale, and for subprocesses 121 and 122 by the heavy-quark mass. \iteme{VINT(54) :} $Q^2$ of the outer hard-scattering subprocess; see {\tt VINT(53)}. \iteme{VINT(55) :} $Q$ scale used as maximum virtuality in parton showers. Is equal to {\tt VINT(53)}, except for deep-inelastic-scattering processes when {\tt MSTP(22)}$ > 0$. \iteme{VINT(56) :} $Q^2$ scale in parton showers; see {\tt VINT(55)}. \iteme{VINT(57) :} $\alphaem$ value of hard process. \iteme{VINT(58) :} $\alphas$ value of hard process. \iteme{VINT(59) :} $\sin\hat{\theta}$ (cf. {\tt VINT(23)}); used for improved numerical precision in elastic and diffractive scattering. \iteme{VINT(61), VINT(62) :} nominal $m^2$ values, i.e. without initial-state radiation effects, for the two partons entering the hard interaction. \iteme{VINT(63), VINT(64) :} nominal $m^2$ values, i.e. without final-state radiation effects, for the two (or one) partons/particles leaving the hard interaction. \iteme{VINT(65) :} $\hat{p}_{\mrm{init}}$, i.e. common nominal absolute momentum of the two partons entering the hard interaction, in their rest frame. \iteme{VINT(66) :} $\hat{p}_{\mrm{fin}}$, i.e. common nominal absolute momentum of the two partons leaving the hard interaction, in their rest frame. \iteme{VINT(67), VINT(68) :} mass of beam and target particle, as {\tt VINT(3)} and {\tt VINT(4)}, except that an incoming $\gamma$ is assigned the $\rho^0$, $\omega$ or $\phi$ mass. Used for elastic scattering $\gamma \p \to \rho^0 \p$ and other similar processes. \iteme{VINT(71) :} $\pTmin$ of process, i.e. {\tt CKIN(3)} or {\tt CKIN(5)}, depending on which is larger, and whether the process is singular in $\pT \to 0$ or not. \iteme{VINT(73) :} $\tau = m^2/s$ value of first resonance, if any; see {\tt MINT(73)}. \iteme{VINT(74) :} $m \Gamma/s$ value of first resonance, if any; see {\tt MINT(73)}. \iteme{VINT(75) :} $\tau = m^2/s$ value of second resonance, if any; see {\tt MINT(74)}. \iteme{VINT(76) :} $m \Gamma/s$ value of second resonance, if any; see {\tt MINT(74)}. \iteme{VINT(80) :} correction factor (evaluated in {\tt PYOFSH}) for the cross section of resonances produced in $2 \to 2$ processes, if only some mass range of the full Breit--Wigner shape is allowed by user-set mass cuts ({\tt CKIN(2)}, {\tt CKIN(45) - CKIN(48)}). \iteme{VINT(81) - VINT(84) :} the $\cos\theta$ and $\varphi$ variables of a true $2 \to 3$ process, where one product is a resonance, effectively giving $2 \to 4$. The first two are $\cos\theta$ and $\varphi$ for the resonance decay, the other two ditto for the effective system formed by the other two particles. \iteme{VINT(85), VINT(86) :} transverse momenta in a true $2 \to 3$ process; one is stored in {\tt VINT(47)} (that of the $\Z^0$ in $\g \g \to \Z^0 \Q \Qbar$), while the smaller of the other two is stored in {\tt VINT(85)} and the larger in {\tt VINT(86)}. \iteme{VINT(91), VINT(92) :} gives a dimensionless suppression factor, to take into account reduction in cross section due to the allowed channels for a $\W^+\W^+$ or $\W^-\W^-$ pair, respectively, in the same sense as {\tt WIDS(24,1)} gives it for a $\W^+\W^-$ pair. \iteme{VINT(98) :} is sum of {\tt VINT(100)} values for current run. \iteme{VINT(99) :} is weight {\tt WTXS} returned from {\tt PYEVWT} call when {\tt MSTP(142)}$\geq 1$, otherwise is 1. \iteme{VINT(100) :} is compensating weight {\tt 1./WTXS} that should be associated with events when {\tt MSTP(142)=1}, otherwise is 1. \iteme{VINT(108) :} ratio of maximum differential cross section observed to maximum differential cross section assumed for the generation; cf. {\tt MSTP(123)}. \iteme{VINT(109) :} ratio of minimal (negative!) cross section observed to maximum differential cross section assumed for the generation; could only become negative if cross sections are incorrectly included. \iteme{VINT(111) - VINT(116) :} for {\tt MINT(61)=1} gives kinematical factors for the different pieces contributing to $\gammaZ$ or $\gamma^*/\Z^0/\Z'^0$ production, for {\tt MINT(61)=2} gives sum of final-state weights for the same; coefficients are given in the order pure $\gamma^*$, $\gamma^*$--$\Z^0$ interference, $\gamma^*$--$\Z'^0$ interference, pure $\Z^0$, $\Z^0$--$\Z'^0$ interference and pure $\Z'^0$. \iteme{VINT(117) :} width of $\Z^0$; needed in $\gamma^*/\Z^0/\Z'^0$ production. \iteme{VINT(121) :} nuclear slope parameter $B$ in $\hat{t}$ distribution for (diffractive and) elastic scattering. \iteme{VINT(122) :} curvature parameter $C$ in $\hat{t}$ distribution for (diffractive and) elastic scattering. \iteme{VINT(131) :} total cross section (in mb) for subprocesses allowed in the pile-up events scenario according to the {\tt MSTP(132)} value. \iteme{VINT(132) :} $\br{n} = ${\tt VINT(131)}$\times${\tt PARP(131)}, cf. {\tt PARI(91)}. \iteme{VINT(133) :} $\langle n \rangle = \sum_i i \, {\cal P}_i / \sum_i {\cal P}_i$ as actually simulated, i.e. $1 \leq i \leq 200$ (or smaller), see {\tt PARI(92)}. \iteme{VINT(134) :} number related to probability to have event in beam--beam crossing; is $\exp(-\br{n}) \sum_i \br{n}^i/i!$ for {\tt MSTP(133)=1} and $\exp(-\br{n}) \sum_i \br{n}^i/(i-1)!$ for {\tt MSTP(133)=2}, cf. {\tt PARI(93)}. \iteme{VINT(138) :} size of the threshold factor (enhancement or suppression) in the latest event with heavy-flavour production; see {\tt MSTP(35)}. \iteme{VINT(141), VINT(142) :} $x$ values for the parton-shower initiators of the hardest interaction; used to find what is left for multiple interactions. \iteme{VINT(143), VINT(144) :} $1 - \sum_i x_i$ for all scatterings; used for rescaling each new $x$-value in the multiple-interaction parton-distribution-function evaluation. \iteme{VINT(145) :} estimate of total parton--parton cross section for multiple interactions; used for {\tt MSTP(82)}$\geq 2$. \iteme{VINT(146) :} common correction factor $f_c$ in the multiple-interaction probability; used for {\tt MSTP(82)}$\geq 2$ (part of $e(b)$, see eq. (\ref{mi:ebenh})). \iteme{VINT(147) :} average hadronic matter overlap; used for {\tt MSTP(82)}$\geq 2$ (needed in evaluation of $e(b)$, see eq. (\ref{mi:ebenh})). \iteme{VINT(148) :} enhancement factor for current event in the multiple-interaction probability, defined as the actual overlap divided by the average one; used for {\tt MSTP(82)}$\geq 2$ (is $e(b)$ of eq. (\ref{mi:ebenh})). \iteme{VINT(149) :} $x_{\perp}^2$ cut-off or turn-off for multiple interactions. For {\tt MSTP(82)}$\leq 1$ it is $4 \pTmin^2/s$, for {\tt MSTP(82)}$\geq 2$ it is $4 p_{\perp 0}^2/s$. \iteme{VINT(150) :} probability to keep the given event in the multiple-interaction scenario, as given by the `Sudakov' form factor. \iteme{VINT(151), VINT(152) :} sum of $x$ values for all the multiple-interaction partons. \iteme{VINT(153) :} current differential cross section value obtained from {\tt PYSIGH}; used in multiple interactions only. \iteme{VINT(155), VINT(156) :} the $x$ value of a photon that branches into quarks or gluons, i.e. $x$ at interface between initial-state QED and QCD cascades. \iteme{VINT(157), VINT(158) :} the primordial $k_{\perp}$ values selected in the two beam remnants. \iteme{VINT(159), VINT(160) :} the $\chi$ values selected for beam remnants that are split into two objects, describing how the energy is shared (see {\tt MSTP(92)} etc.); is 0 if no splitting is needed. \iteme{VINT(161) - VINT(200) :} sum of Cabibbo--Kobayashi--Maskawa squared matrix elements that a given flavour is allowed to couple to. Results are stored in format {\tt VINT(180+KF)} for quark and lepton flavours and antiflavours (which need not be the same; see {\tt MDME(IDC,2)}). For leptons, these factors are normally unity. \iteme{VINT(201) - VINT(220) :} additional variables needed in phase-space selection for $2 \to 3$ processes with {\tt ISET(ISUB)=5}. Below indices 1, 2 and 3 refer to scattered partons 1, 2 and 3, except that the $q$ four-momentum variables are $q_1 + q_2 \to q_1' q_2' q_3'$. All kinematical variables refer to the internal kinematics of the 3-body final state --- the kinematics of the system as a whole is described by $\tau'$ and $y$, and the mass distribution of particle 3 (a resonance) by $\tau$. \begin{subentry} \iteme{VINT(201) :} $m_1$. \iteme{VINT(202) :} $p_{\perp 1}^2$. \iteme{VINT(203) :} $\varphi_1$. \iteme{VINT(204) :} $M_1$ (mass of propagator particle). \iteme{VINT(205) :} weight for the $p_{\perp 1}^2$ choice. \iteme{VINT(206) :} $m_2$. \iteme{VINT(207) :} $p_{\perp 2}^2$. \iteme{VINT(208) :} $\varphi_2$. \iteme{VINT(209) :} $M_2$ (mass of propagator particle). \iteme{VINT(210) :} weight for the $p_{\perp 2}^2$ choice. \iteme{VINT(211) :} $y_3$. \iteme{VINT(212) :} $y_{3 max}$. \iteme{VINT(213) :} $\epsilon = \pm 1$; choice between two mirror solutions $1 \leftrightarrow 2$. \iteme{VINT(214) :} weight associated to $\epsilon$-choice. \iteme{VINT(215) :} $t_1 = (q_1 - q_1')^2$. \iteme{VINT(216) :} $t_2 = (q_2 - q_2')^2$. \iteme{VINT(217) :} $q_1 q_2'$ four-product. \iteme{VINT(218) :} $q_2 q_1'$ four-product. \iteme{VINT(219) :} $q_1' q_2'$ four-product. \iteme{VINT(220) :} $\sqrt{(m_{\perp 12}^2 - m_{\perp 1}^2 - m_{\perp 2}^2)^2 - 4 m_{\perp 1}^2 m_{\perp 2}^2}$, where $m_{\perp 12}$ is the transverse mass of the $q'_1 q'_2$ system. \end{subentry} \iteme{VINT(221) - VINT(225) :} $\theta$, $\varphi$ and \mbox{{\boldmath $\beta$}} for rotation and boost from c.m. frame to hadronic c.m. frame of a lepton--hadron event. \iteme{VINT(231) :} $Q^2_{\mmin}$ scale for current parton-distribution function set. \iteme{VINT(281) :} for resolved photon events, it gives the ratio between the total $\gamma X$ cross section and the total $\pi^0 X$ cross section, where $X$ represents the target particle. \iteme{VINT(282) :} virtuality scale at which an anomalous photon is resolved or, more precisely, the $\pT^2$ of the $\gamma \to \q\qbar$ vertex. \iteme{VINT(289) :} squared c.m. energy found in {\tt PYINIT} call. \iteme{VINT(290) :} the {\tt WIN} argument of a {\tt PYINIT} call. \iteme{VINT(291) - VINT(300) :} the two five-vectors of the two incoming particles, as reconstructed in {\tt PYINKI} or {\tt PYEVKI}. These may vary from one event to the next. \end{entry} \drawbox{COMMON/PYINT2/ISET(200),KFPR(200,2),COEF(200,20),ICOL(40,4,2)}% \label{p:PYINT2} \begin{entry} \itemc{Purpose:} to store information necessary for efficient generation of the different subprocesses, specifically type of generation scheme and coefficients of the Jacobian. Also to store allowed colour-flow configurations. These variables must not be changed by you. \iteme{ISET(ISUB) :}\label{p:ISET} gives the type of kinematical-variable selection scheme used for subprocess ISUB. \begin{subentry} \iteme{= 0 :} elastic, diffractive and low-$\pT$ processes. \iteme{= 1 :} $2 \to 1$ processes (irrespective of subsequent decays). \iteme{= 2 :} $2 \to 2$ processes (i.e. the bulk of processes). \iteme{= 3 :} $2 \to 3$ processes (like $\q \q' \to \q'' \q''' \H^0$). \iteme{= 4 :} $2 \to 4$ processes (like $\q \q' \to \q'' \q''' \W^+ \W^-$). \iteme{= 5 :} `true' $2 \to 3$ processes, one method. \iteme{= 6 :} `true' $2 \to 3$ processes, another method; currently only $\g \g \to \Z^0 \Q \Qbar$. \iteme{= 9 :} $2 \to 2$ in multiple interactions ($\pT$ as kinematics variable). \iteme{= 11 :} a user-defined process. \iteme{= -1 :} legitimate process which has not yet been implemented. \iteme{= -2 :} ISUB is an undefined process code. \end{subentry} \iteme{KFPR(ISUB,J) :}\label{p:KFPR} give the KF flavour codes for the products produced in subprocess ISUB. If there is only one product, the {\tt J=2} position is left blank. Also, quarks and leptons assumed massless in the matrix elements are denoted by 0. The main application is thus to identify resonances produced ($\Z^0$, $\W^{\pm}$, $\H^0$, etc.). \iteme{COEF(ISUB,J) :}\label{p:COEF} factors used in the Jacobians in order to speed up the selection of kinematical variables. More precisely, the shape of the cross section is given as the sum of terms with different behaviour, where the integral over the allowed phase space is unity for each term. {\tt COEF} gives the relative strength of these terms, normalized so that the sum of coefficients for each variable used is unity. Note that which coefficients are indeed used is process-dependent. \begin{subentry} \iteme{ISUB :} standard subprocess code. \iteme{J = 1 :} $\tau$ selected according $1/\tau$. \iteme{J = 2 :} $\tau$ selected according to $1/\tau^2$. \iteme{J = 3 :} $\tau$ selected according to $1/(\tau(\tau+\tau_R))$, where $\tau_R = m_R^2/s$ is $\tau$ value of resonance; only used for resonance production. \iteme{J = 4 :} $\tau$ selected according to Breit--Wigner of form $1/((\tau-\tau_R)^2+\gamma_R^2)$, where $\tau_R = m_R^2/s$ is $\tau$ value of resonance and $\gamma_R = m_R \Gamma_R/s$ is its scaled mass times width; only used for resonance production. \iteme{J = 5 :} $\tau$ selected according to $1/(\tau(\tau+\tau_{R'}))$, where $\tau_{R'} = m_{R'}^2/s$ is $\tau$ value of second resonance; only used for simultaneous production of two resonances. \iteme{J = 6 :} $\tau$ selected according to second Breit--Wigner of form $1/((\tau-\tau_{R'})^2+\gamma_{R'}^2)$, where $\tau_{R'} = m_{R'}^2/s$ is $\tau$ value of second resonance and $\gamma_{R'} = m_{R'} \Gamma_{R'}/s$ is its scaled mass times width; is used only for simultaneous production of two resonances, like $\gamma^*/\Z^0/\Z'^0$. \iteme{J = 7 :} $\tau$ selected according to $1/(1-\tau)$; only used when both parton distributions are peaked at $x = 1$. \iteme{J = 8 :} $y$ selected according to $y - y_{\mmin}$. \iteme{J = 9 :} $y$ selected according to $y_{\mmax} - y$. \iteme{J = 10 :} $y$ selected according to $1/\cosh(y)$. \iteme{J = 11 :} $y$ selected according to $1/(1-\exp(y-y_{\mmax}))$; only used when beam parton distribution is peaked close to $x = 1$. \iteme{J = 12 :} $y$ selected according to $1/(1-\exp(y_{\mmin}-y))$; only used when target parton distribution is peaked close to $x = 1$. \iteme{J = 13 :} $z = \cos\hat{\theta}$ selected evenly between limits. \iteme{J = 14 :} $z = \cos\hat{\theta}$ selected according to $1/(a-z)$, where $a = 1 + 2 m_3^2 m_4^2/\hat{s}^2$, $m_3$ and $m_4$ being the masses of the two final-state particles. \iteme{J = 15 :} $z = \cos\hat{\theta}$ selected according to $1/(a+z)$, with $a$ as above. \iteme{J = 16 :} $z = \cos\hat{\theta}$ selected according to $1/(a-z)^2$, with $a$ as above. \iteme{J = 17 :} $z = \cos\hat{\theta}$ selected according to $1/(a+z)^2$, with $a$ as above. \iteme{J = 18 :} $\tau'$ selected according to $1/\tau'$. \iteme{J = 19 :} $\tau'$ selected according to $(1 - \tau/\tau')^3/\tau'^2$. \iteme{J = 20 :} $\tau'$ selected according to $1/(1-\tau')$; only used when both parton distributions are peaked close to $x = 1$. \end{subentry} \iteme{ICOL :}\label{p:ICOL} contains information on different colour-flow topologies in hard $2 \to 2$ processes. \end{entry} \drawbox{COMMON/PYINT3/XSFX(2,-40:40),ISIG(1000,3),SIGH(1000)}% \label{p:PYINT3} \begin{entry} \itemc{Purpose:} to store information on parton distributions, subprocess cross sections and different final-state relative weights. These variables must not be changed by you. \iteme{XSFX :}\label{p:XSFX} current values of parton-distribution functions $xf(x)$ on beam and target side. \iteme{ISIG(ICHN,1) :}\label{p:ISIG} incoming parton/particle on the beam side to the hard interaction for allowed channel number {\tt ICHN}. The number of channels filled with relevant information is given by {\tt NCHN}, one of the arguments returned in a {\tt PYSIGH} call. Thus only $1 \leq ${\tt ICHN}$ \leq ${\tt NCHN} is filled with relevant information. \iteme{ISIG(ICHN,2) :} incoming parton/particle on the target side to the hard interaction for allowed channel number {\tt ICHN}. See also comment above. \iteme{ISIG(ICHN,3) :} colour-flow type for allowed channel number {\tt ICHN}; see {\tt MSTI(2)} list. See also above comment. For `subprocess' 96 uniquely, {\tt ISIG(ICHN,3)} is also used to translate information on what is the correct subprocess number (11, 12, 13, 28, 53 or 68); this is used for reassigning subprocess 96 to either of these. \iteme{SIGH(ICHN) :}\label{p:SIGH} evaluated differential cross section for allowed channel number {\tt ICHN}, i.e. matrix-element value times parton distributions, for current kinematical setup (in addition, Jacobian factors are included in the figures, as used to speed up generation). See also comment for {\tt ISIG(ICHN,1)}. \end{entry} \drawbox{COMMON/PYINT4/WIDP(21:40,0:40),WIDE(21:40,0:40),WIDS(21:40,3)}% \label{p:PYINT4} \begin{entry} \itemc{Purpose:} to store partial and effective decay widths for the different resonances. These variables must not be changed by you. \iteme{WIDP(KF,J) :}\label{p:WIDP} gives partial decay widths of resonances into different channels (in GeV), given that all physically allowed final states are included. \begin{subentry} \iteme{KF :} standard KF code for resonance considered. When top is treated like a resonance (see {\tt MSTP(48)}) it is stored in position 26. When the fourth generation fermions $\lrm$, $\hrm$, $\chi^-$ and $\nu_{\chi}$ are treated like resonances (see {\tt MSTP(49)}) they are stored in positions 27, 28, 29 and 30, respectively. \iteme{J :} enumerates the different decay channels possible for resonance KF, as stored in the {\Je} {\tt LUDAT3} common block, with the first channel in {\tt J=1}, etc. \end{subentry} \iteme{WIDE(KF,J) :}\label{p:WIDE} gives effective decay widths of resonances into different channels (in GeV), given the decay modes actually left open in the current run. The on/off status of decay modes is set by the {\tt MDME} switches in {\Je}; see section \ref{ss:parapartdat}. \begin{subentry} \iteme{KF :} standard KF code for resonance considered. For comment about top and fourth generation see {\tt WIDP} above. \iteme{J :} enumerates the different decay channels possible for resonance KF, as stored in the {\Je} {\tt LUDAT3} common block, with the first channel in {\tt J=1}, etc. \end{subentry} \iteme{WIDS(KF,J) :}\label{p:WIDS} gives a dimensionless suppression factor, which is defined as the ratio of the total width of channels switched on to the total width of all possible channels (replace width by squared width for a pair of resonances). The on/off status of channels is set by the {\tt MDME} switches in {\Je}; see section \ref{ss:parapartdat}. The information in {\tt WIDS} is used e.g. in cross-section calculations. \begin{subentry} \iteme{KF :} standard KF code for resonance considered. For comment about top and fourth generation see {\tt WIDP} above. \iteme{J = 1 :} suppression when a pair of resonances of type KF are produced together. When an antiparticle exists, the particle--antiparticle pair (such as $\W^+ \W^-$) is the relevant combination, else the particle--particle one (such as $\Z^0 \Z^0$). \iteme{J = 2 :} suppression for a particle of type KF when produced on its own, or together with a particle of another type. \iteme{J = 3 :} suppression for an antiparticle of type KF when produced on its own, or together with a particle of another type. \end{subentry} \end{entry} \drawbox{COMMON/PYINT5/NGEN(0:200,3),XSEC(0:200,3)}\label{p:PYINT5} \begin{entry} \itemc{Purpose:} to store information necessary for cross-section calculation and differential cross-section maximum violation. These variables must not be changed by you. \iteme{NGEN(ISUB,1) :}\label{p:NGEN} gives the number of times that the differential cross section (times Jacobian factors) has been evaluated for subprocess ISUB, with {\tt NGEN(0,1)} the sum of these. \iteme{NGEN(ISUB,2) :} gives the number of times that a kinematical setup for subprocess ISUB is accepted in the generation procedure, with {\tt NGEN(0,2)} the sum of these. \iteme{NGEN(ISUB,3) :} gives the number of times an event of subprocess type ISUB is generated, with {\tt NGEN(0,3)} the sum of these. Usually {\tt NGEN(ISUB,3) = NGEN(ISUB,2)}, i.e. an accepted kinematical configuration can normally be used to produce an event. \iteme{XSEC(ISUB,1) :}\label{p:XSEC} estimated maximum differential cross section (times the Jacobian factors used to speed up the generation process) for the different subprocesses in use, with {\tt XSEC(0,1)} the sum of these (except low-$\pT$, i.e. ISUB = 95). \iteme{XSEC(ISUB,2) :} gives the sum of differential cross sections (times Jacobian factors) for the {\tt NGEN(ISUB,1)} phase-space points evaluated so far. \iteme{XSEC(ISUB,3) :} gives the estimated integrated cross section for subprocess ISUB, based on the statistics accumulated so far, with {\tt XSEC(0,3)} the estimated total cross section for all subprocesses included (all in mb). This is exactly the information obtainable by a {\tt PYSTAT(1)} call. \end{entry} \drawboxtwo{COMMON/PYINT6/PROC(0:200)}{CHARACTER PROC*28}% \label{p:PYINT6} \begin{entry} \itemc{Purpose:} to store character strings for the different possible subprocesses; used when printing tables. \iteme{PROC(ISUB) :}\label{p:PROC} name for the different subprocesses, according to ISUB code. {\tt PROC(0)} denotes all processes. \end{entry} \drawbox{COMMON/PYINT7/SIGT(0:6,0:6,0:5)}\label{p:PYINT7} \begin{entry} \itemc{Purpose:} to store information on total, elastic and diffractive cross sections. These variables should only be set by you for the option {\tt MSTP(31)=0}; else they should not be touched. All numbers are given in mb. \iteme{SIGT(I1,I2,J) :}\label{p:SIGT} the cross section, both total and subdivided by class (elastic, diffractive etc.). For a photon to be considered as a VMD meson the cross sections are additionally split into the contributions from the various meson states. \begin{subentry} \iteme{I1, I2 :} allowed states for the incoming particle on side 1 and 2, respectively. \begin{subentry} \iteme{= 0 :} sum of all allowed states. Except for a photon to be considered as a VMD meson this is the only nonvanishing entry. \iteme{= 1 :} the contribution from the $\rho^0$ VMD state. \iteme{= 2 :} the contribution from the $\omega$ VMD state. \iteme{= 3 :} the contribution from the $\phi$ VMD state. \iteme{= 4 :} the contribution from the $\Jpsi$ VMD state. \iteme{= 5, 6 :} reserved for future use. \end{subentry} \iteme{J :} the total and partial cross sections. \begin{subentry} \iteme{= 0 :} the total cross section. \iteme{= 1 :} the elastic cross section. \iteme{= 2 :} the single diffractive cross section $AB \to XB$. \iteme{= 3 :} the single diffractive cross section $AB \to AX$. \iteme{= 4 :} the double diffractive cross section. \iteme{= 5 :} the inelastic, non-diffractive cross section. \end{subentry} \itemc{Warning:} If you set these values yourself, it is important that they are internally consistent, since this is not explicitly checked by the program. Thus the contributions {\tt J=}1--5 should add up to the {\tt J=}0 one and, for VMD photons, the contributions {\tt I=}1--4 should add up to the {tt I=}0 one. \end{subentry} \end{entry} \boxsep Finally, in addition a number of routines and common blocks with names beginning with {\tt RK} come with the program. These contain the matrix-element evaluation for the process $\g \g \to \Z \q \qbar$, based on a program of Ronald Kleiss, with only minor modifications. \subsection{Examples} The program is built as a slave system, i.e. you supply the main program, which calls on the {\Py} and {\Je} routines to perform specific tasks and then resumes control. A typical program for the analysis of collider events at 630 GeV c.m. energy with a minimum $\pT$ of 10 GeV/c at the hard scattering (because of initial-state radiation, fragmentation effects, etc., the actual $\pT$ cut-off will be smeared around this value) might look like \begin{verbatim} COMMON/LUJETS/N,K(4000,5),P(4000,5),V(4000,5) COMMON/PYSUBS/MSEL,MSUB(200),KFIN(2,-40:40),CKIN(200) COMMON/PYPARS/MSTP(200),PARP(200),MSTI(200),PARI(200) ... ! set all common block variables that ... ! did not have desired default values CKIN(3)=10. ! lower p_T cut-off CALL PYINIT('CMS','p','pbar',630.) ! initialize ... ! initialize analysis statistics DO 100 IEVENT=1,1000 ! loop over events CALL PYEVNT ! generate event IF(IEVENT.EQ.1) CALL LULIST(1) ! list first event ... ! insert desired analysis chain for ... ! each event 100 CONTINUE CALL PYSTAT(1) ! print cross sections ... ! user output END \end{verbatim} \clearpage \section{Initial- and Final-State Radiation} \label{s:showinfi} Starting from the hard interaction, initial- and final-state radiation corrections may be added. This is normally done by making use of the parton-shower language --- only for the $\ee \to \q \qbar$ process does {\Je} offer a matrix-element option (described in section \ref{ss:eematrix}). The algorithms used to generate initial- and final-state showers are rather different, and are therefore described separately below, starting with the conceptually easier final-state one. Before that, some common elements are introduced. The main reference for final-state showers is ref. \cite{Ben87a} and for initial-state ones ref. \cite{Sjo85}. \subsection{Shower Evolution} In the leading log picture, a shower may be viewed as a sequence of $1 \to 2$ branchings $a \to bc$. Here $a$ is called the mother and $b$ and $c$ the two daughters. Each daughter is free to branch in its turn, so that a tree-like stucture can evolve. We will use the work `parton' for all the objects $a$, $b$ and $c$ involved in the branching process, i.e. not only for quarks and gluons but also for leptons and photons. The branchings included in the program are $\q \to \q \g$, $\g \to \g \g$, $\g \to \q \qbar$, $\q \to \q \gamma$ and $\ell \to \ell \gamma$. Photon branchings, i.e. $\gamma \to \q \qbar$ and $\gamma \to \ell \br{\ell}$, have not been included so far, since they are reasonably rare and since no urgent need for them has been perceived. \subsubsection{The evolution equations} In the shower formulation, the kinematics of each branching is given in terms of two variables, $Q^2$ and $z$. Slightly different interpretations may be given to these variables, and indeed this is one main area where the various programs on the market differ. $Q^2$ has dimensions of squared mass, and is related to the mass or transverse momentum scale of the branching. $z$ gives the sharing of the $a$ energy and momentum between the two daughters, with parton $b$ taking a fraction $z$ and parton $c$ a fraction $1-z$. To specify the kinematics, an azimuthal angle $\varphi$ of the $b$ around the $a$ direction is needed in addition; normally $\varphi$ is chosen to be isotropically distributed, although options for non-isotropic distributions exist. The probability for a parton to branch is given by the evolution equations (also called DGLAP or Altarelli--Parisi \cite{Gri72,Alt77}). It is convenient to introduce \begin{equation} t = \ln(Q^2/\Lambda^2) ~~~ \Rightarrow ~~~ \d t = \d \ln(Q^2) = \frac{\d Q^2}{Q^2} ~, \end{equation} where $\Lambda$ is the QCD $\Lambda$ scale in $\alphas$. Of course, this choice is more directed towards the QCD parts of the shower, but it can be used just as well for the QED ones. In terms of the two variables $t$ and $z$, the differential probability $\d {\cal P}$ for parton $a$ to branch is now \begin{equation} \d {\cal P}_a = \sum_{b,c} \frac{\alpha_{abc}}{2 \pi} \, P_{a \to bc}(z) \, \d t \, \d z ~. \label{sh:Patobc} \end{equation} Here the sum is supposed to run over all allowed branchings, for a quark $\q \to \q \g$ and $\q \to \q \gamma$, and so on. The $\alpha_{abc}$ factor is $\alphaem$ for QED branchings and $\alphas$ for QCD ones (to be evaluated at some suitable scale, see below). The splitting kernels $P_{a \to bc}(z)$ are \begin{eqnarray} P_{\q \to \q\g}(z)&=&C_F \, \frac{1+z^2}{1-z} ~, \nonumber \\ P_{\g \to \g\g}(z)&=&N_C \, \frac{(1-z(1-z))^2}{z(1-z)} ~, \nonumber \\ P_{\g \to \q\qbar}(z)&=&T_R \, (z^2 + (1-z)^2) ~, \nonumber \\ P_{\q \to \q\gamma}(z)&=&e_{\q}^2 \, \frac{1+z^2}{1-z} ~, \nonumber \\ P_{\ell \to \ell\gamma}(z)&=&e_{\ell}^2 \, \frac{1+z^2}{1-z} ~, \label{sh:APker} \end{eqnarray} with $C_F = 4/3$, $N_C = 3$, $T_R = n_f/2$ (i.e. $T_R$ receives a contribution of $1/2$ for each allowed $\q\qbar$ flavour), and $e_{\q}^2$ and $e_{\ell}^2$ the squared electric charge ($4/9$ for $\u$-type quarks, $1/9$ for $\d$-type ones, and 1 for leptons). Persons familiar with analytical calculations may wonder why the `+ prescriptions' and $\delta(1-z)$ terms of the splitting kernels in eq. (\ref{sh:APker}) are missing. These complications fulfil the task of ensuring flavour and energy conservation in the analytical equations. The corresponding problem is solved trivially in Monte Carlo programs, where the shower evolution is traced in detail, and flavour and four-momentum are conserved at each branching. The legacy left is the need to introduce a cut-off on the allowed range of $z$ in splittings, so as to avoid the singular regions corresponding to excessive production of very soft gluons. Also note that $P_{\g \to \g\g}(z)$ is given here with a factor $N_C$ in front, while it is sometimes shown with $2 N_C$. The confusion arises because the final state contains two identical partons. With the normalization above, $P_{a \to bc}(z)$ is interpreted as the branching probability for the original parton $a$. On the other hand, one could also write down the probability that a parton $b$ is produced with a fractional energy $z$. Almost all the above kernels can be used unchanged also for this purpose, with the obvious symmetry $P_{a \to bc}(z) = P_{a \to cb}(1-z)$. For $\g \to \g\g$, however, the total probability to find a gluon with energy fraction $z$ is the sum of the probability to find either the first or the second daughter there, and that gives the factor of 2 enhancement. \subsubsection{The Sudakov form factor} \label{ss:sudakov} The $t$ variable fills the function of a kind of time for the shower evolution. In final-state showers, $t$ is constrained to be gradually decreasing away from the hard scattering, in initial-state ones to be gradually increasing towards the hard scattering. This does not mean that an individual parton runs through a range of $t$ values: in the end, each parton is associated with a fixed $t$ value, and the evolution procedure is just a way of picking that value. It is only the ensemble of partons in many events that evolves continuously with $t$, cf. the concept of parton distributions. For a given $t$ value we define the integral of the branching probability over all allowed $z$ values, \begin{equation} {\cal I}_{a \to bc}(t) = \int_{z_{-}(t)}^{z_{+}(t)} \d z \, \frac{\alpha_{abc}}{2 \pi} \, P_{a \to bc}(z) ~. \end{equation} The na\"{\i}ve probability that a branching occurs during a small range of $t$ values, $\delta t$, is given by $\sum_{b,c} {\cal I}_{a \to bc}(t) \, \delta t$, and thus the probability for no emission by $1 - \sum_{b,c} {\cal I}_{a \to bc}(t) \, \delta t$. If the evolution of parton $a$ starts at a `time' $t_0$, the probability that the parton has not yet branched at a `later time' $t > t_0$ is given by the product of the probabilities that it did not branch in any of the small intervals $\delta t$ between $t_0$ and $t$. In other words, letting $\delta t \to 0$, the no-branching probability exponentiates: \begin{equation} {\cal P}_{\mrm{no-branching}}(t_0,t) = \exp \left\{ - \int_{t_0}^t \d t' \, \sum_{b,c} {\cal I}_{a \to bc}(t') \right\} = S_a(t) ~. \label{sh:Pnobranch} \end{equation} Thus the actual probability that a branching of $a$ occurs at $t$ is given by \begin{equation} \frac{\d {\cal P}_a}{\d t} = - \frac{\d {\cal P}_{\mrm{no-branching}}(t_0,t)}{\d t} = \left( \sum_{b,c} {\cal I}_{a \to bc}(t) \right) \exp \left\{ - \int_{t_0}^t \d t' \, \sum_{b,c} {\cal I}_{a \to bc}(t') \right\} ~. \label{sh:Pbranch} \end{equation} The first factor is the na\"{\i}ve branching probability, the second the suppression due to the conservation of total probability: if a parton has already branched at a `time' $t' < t$, it can no longer branch at $t$. This is nothing but the exponential factor that is familiar from radioactive decay. In parton-shower language the exponential factor $S_a(t) = {\cal P}_{\mrm{no-branching}}(t_0,t)$ is referred to as the Sudakov form factor \cite{Sud56}. The ordering in terms of increasing $t$ above is the appropriate one for initial-state showers. In final-state showers the evolution is from an initial $t_{\mmax}$ (set by the hard scattering) and towards smaller $t$. In that case the integral from $t_0$ to $t$ in eqs. (\ref{sh:Pnobranch}) and (\ref{sh:Pbranch}) is replaced by an integral from $t$ to $t_{\mmax}$. Since, by convention, the Sudakov factor is still defined from the lower cut-off $t_0$, i.e. gives the probability that a parton starting at scale $t$ will not have branched by the lower cut-off scale $t_0$, the no-branching factor is actually ${\cal P}_{\mrm{no-branching}}(t_{\mmax},t) = S_a(t_{\mmax})/S_a(t)$. We note that the above structure is exactly of the kind discussed in section \ref{ss:vetoalg}. The veto algorithm is therefore extensively used in the Monte Carlo simulation of parton showers. \subsubsection{Matching to the hard scattering} \label{sss:showermatching} The evolution in $Q^2$ is begun from some maximum scale $Q_{\mmax}^2$ for final-state parton showers, and is terminated at (a possibly different) $Q_{\mmax}^2$ for initial-state showers. In general $Q_{\mmax}^2$ is not known. Indeed, since the parton-shower language does not guarantee agreement with higher-order matrix-element results, neither in absolute shape nor normalization, there is no unique prescription for a `best' choice. Generically $Q_{\mmax}$ should be of the order of the hard-scattering scale, i.e. the largest virtuality should be associated with the hard scattering, and initial- and final-state parton showers should only involve virtualities smaller than that. This may be viewed just as a matter of sound bookkeeping: in a $2 \to n$ graph, a $2 \to 2$ hard-scattering subgraph could be chosen several different ways, but if all the possibilities were to be generated then the cross section would be double-counted. Therefore one should define the $2 \to 2$ `hard' piece of a $2 \to n$ graph as the one that involves the largest virtuality. Of course, the issue of double-counting depends a bit on what processes are actually generated in the program. If one considers a $\q \qbar \g$ final state in hadron colliders, this could come either as final-state radiation off a $\q \qbar$ pair, or by a gluon splitting in a $\q \qbar$ pair, or many other ways, so that the danger of double-counting is very real. On the other hand, consider the production of a low-$\pT$, low-mass Drell--Yan pair of leptons, together with two quark jets. Such a process in principle could proceed by having a $\gamma^*$ emitted off a quark leg, with a quark--quark scattering as hard interaction. However, since this process is not included in the program, there is no actual danger of (this particular) double-counting, and so the scale of evolution could be picked larger than the mass of the Drell--Yan pair, at least by some amount. For most $2 \to 2$ scattering processes in {\Py}, the $Q^2$ scale of the hard scattering is chosen to be $Q_{\mrm{hard}}^2 = \pT^2$ (when the final-state particles are massless, otherwise masses are added). In final-state showers, where $Q$ is associated with the mass of the branching parton, transverse momenta generated in the shower are constrained by $\pT < Q/2$. An ordering that shower $\pT$ be smaller than the hard-scattering $\pT$ therefore corresponds roughly to $Q_{\mmax}^2 = 4 Q_{\mrm{hard}}^2$, which is the default assumption. In principle, the constraints are slightly different for initial-state showers, but not enough to warrant a separate $Q_{\mmax}$ choice. The situation is rather better for the final-state showers in the decay of any colour-singlet particles, such as the $\Z^0$ or the $\H^0$, either as part of a hard $2 \to 1 \to 2$ process, or anywhere else in the final state. Then we know that $Q_{\mmax}$ has to be put equal to the particle mass. It is also possible to match the parton-shower evolution to the first-order matrix-element results. In the program this is done under the assumption that the resonance has spin one, and this approach is known to work very well for $\gammaZ$. The machinery is not fully correct for the spin-zero $\H^0$, but should also there provide a rather good description. QCD processes such as $\q\g \to \q\g$ pose a special problem when the scattering angle is small. Coherence effects (see below) may then restrict the emission further than what is just given by the $Q_{\mmax}$ scale introduced above. This is most easily viewed in the rest frame of the $2 \to 2$ hard scattering subprocess. Some colours flow from the initial to the final state. The radiation associated with such a colour flow should be restricted to a cone with opening angle given by the difference between the original and the final colour directions; there is one such cone around the incoming parton for initial state radiation and one around the outgoing parton for final state radiation. Colours that are annihilated or created in the process effectively correspond to an opening angle of 180$^{\circ}$ and therefore the emission is not constrained for these. For a gluon, which have two colours and therefore two different cones, a random choice is made between the two for the first branching. Further, coherence effects also implie azimuthal anisotropies of the emission inside the allowed cones. \subsection{Final-State Showers} Final-state showers are time-like, i.e. all virtualities $m^2 = E^2 - \mbf{p}^2 \geq 0$. The maximum allowed virtuality scale $Q^2_{\mmax}$ is set by the hard-scattering process, and thereafter the virtuality is decreased in each subsequent branching, down to the cut-off scale $Q_0^2$. This cut-off scale is used to regulate both soft and collinear divergences in the emission probabilities. The main points of the {\Je} showering algorithm are as follows. \begin{Itemize} \item It is a leading-log algorithm, of the improved, coherent kind, i.e. with angular ordering. \item It can be used for an arbitrary initial pair of partons or, in fact, for any one, two or three given entities (including hadrons and gauge bosons) although only quarks, gluons and leptons can initiate a shower. \item The pair of showering partons may be given in any frame, but the evolution is carried out in the c.m. frame of the showering partons. \item Energy and momentum are conserved exactly at each step of the showering process. \item If the initial pair is $\q\qbar$ or $\ell^+ \ell^-$ (coming from a resonance decay) an additional rejection technique is used in the first branching of each of the two original partons, so as to reproduce the lowest-order differential 3-jet cross section. \item In subsequent branchings, angular ordering (coherence effects) is imposed. \item Gluon helicity effects, i.e. correlations between the production plane and the decay plane of a gluon, can be included. \item The first-order $\alphas$ expression is used, with the $Q^2$ scale given by (an approximation to) the squared transverse momentum of a branching. The default $\Lambda_{\mrm{QCD}}$, which should not be regarded as a proper $\Lambda_{\br{\mrm{MS}}}$, is 0.4 GeV. \item The parton shower is by default cut off at a mass scale of 1 GeV. \end{Itemize} Let us now proceed with a more detailed description. \subsubsection{The choice of evolution variable} In the {\Je} shower algorithm, the evolution variable $Q^2$ is associated with the squared mass of the branching parton, $Q^2 = m_a^2$ for a branching $a \to bc$. As a consequence, $t = \ln(Q^2/\Lambda^2) = \ln(m_a^2/\Lambda^2)$. This $Q^2$ choice is not unique, and indeed other programs have other definitions: {\sc Herwig} uses $Q^2 \approx m^2/(2z(1-z))$ \cite{Mar88} and {\sc Ariadne} $Q^2 = \pT^2 \approx z(1-z)m^2$ \cite{Pet88}. With $Q$ a mass scale, the lower cut-off $Q_0$ is one in mass. To be more precise, in a QCD shower, the $Q_0$ parameter is used to derive effective masses \begin{eqnarray} m_{\mrm{eff},\g} & = & \frac{1}{2} Q_0 ~, \nonumber \\ m_{\mrm{eff},\q} & = & \sqrt{ m_{\q}^2 + \frac{1}{4} Q_0^2 } ~, \end{eqnarray} where the $m_{\q}$ have been chosen as typical current-algebra quark masses. A parton cannot branch unless its mass is at least the sum of the lightest pair of allowed decay products, i.e. the minimum mass scale at which a branching is possible is \begin{eqnarray} m_{\mmin,\g} & = & 2 \, m_{\mrm{eff},\g} = Q_0 ~, \nonumber \\ m_{\mmin,\q} & = & m_{\mrm{eff},\q} + m_{\mrm{eff},\g} \geq Q_0 ~. \end{eqnarray} The above masses are used to constrain the allowed range of $Q^2$ and $z$ values. However, once it has been decided that a parton cannot branch any further, that parton is put on the mass shell, i.e. `final-state' gluons are massless. When also photon emission is included, a separate $Q_0$ scale is introduced for the QED part of the shower, exactly reproducing the QCD one above \cite{Sjo92c}. By default the two $Q_0$ scales are chosen equal, and have the value 1 GeV. If anything, one would be inclined to allow a cut-off lower for photon emission than for gluon one. In that case the allowed $z$ range of photon emission would be larger than that of gluon emission, and at the end of the shower evolution only photon emission would be allowed. Photon and gluon emission differ fundamentally in that photons appear as physical particles in the final state, while gluons are confined. For photon emission off quarks, however, the confinement forces acting on the quark may provide an effective photon emission cut-off at larger scales than the bare quark mass. Soft and collinear photons could also be emitted by the final-state charged hadrons; the matching between emission off quarks and off hadrons is a delicate issue, and we therefore do not attempt to address the soft-photon region. For photon emission off leptons, there is no need to introduce any collinear emission cut-off beyond what is given by the lepton mass, but we keep the same cut-off approach as for quarks: firstly, the program is not aimed at high-precision studies of lepton pairs (where interference terms between initial- and final-state radiation also would have to be included); secondly, most experimental procedures would include the energy of collinear photons into the effective energy of a final-state lepton. \subsubsection{The choice of energy splitting variable} The final-state radiation machinery is always applied in the c.m. frame of the hard scattering, from which normally emerges a pair of evolving partons. Occasionally there may be one evolving parton recoiling against a non-evolving one, as in $\q \qbar \to \g \gamma$, where only the gluon evolves in the final state, but where the energy of the photon is modifed by the branching activity of the gluon. (With only one evolving parton and nothing else, it would not be possible to conserve energy and momentum when the parton is assigned a mass.) Thus, before the evolution is performed, the parton pair is boosted to their common c.m. frame, and rotated to sit along the $z$ axis. After the evolution, the full parton shower is rotated and boosted back to the original frame of the parton pair. The interpretation of the energy and momentum splitting variable $z$ is not unique, and in fact the program allows the possibility to switch between four different alternatives \cite{Ben87a}, `local' and `global' $z$ definition combined with `constrained' or `unconstrained' evolution. In all four of them, the $z$ variable is interpreted as an energy fraction, i.e. $E_b = z E_a$ and $E_c = (1-z) E_a$. In the `local' choice of $z$ definition, energy fractions are defined in the rest frame of the grandmother, i.e. the mother of parton $a$. The preferred choice is the `global' one, in which energies are always evaluated in the c.m. frame of the hard scattering. The two definitions agree for the branchings of the partons that emerge directly from the hard scattering, since the hard scattering itself is considered to be the `mother' of the first generation of partons. For instance, in $\Z^0 \to \q\qbar$ the $\Z^0$ is considered the mother of the $\q$ and $\qbar$, even though the branching is not handled by the parton-showering machinery. The `local' and `global' definitions diverge for subsequent branchings, where the `global' tends to allow more shower evolution. In a branching $a \to bc$ the kinematically allowed range of $z = z_a$ values, $z_{-} < z < z_{+}$, is given by \begin{equation} z_{\pm} = \frac{1}{2} \left\{ 1 + \frac{m_b^2 - m_c^2}{m_a^2} \pm \frac{|\mbf{p}_a|}{E_a} \, \frac{\sqrt{ (m_a^2 - m_b^2 - m_c^2)^2 - 4 m_b^2 m_c^2}}{m_a^2} \right\} ~. \label{sh:zbounds} \end{equation} With `constrained' evolution, these bounds are respected in the evolution. The cut-off masses $m_{\mrm{eff},b}$ and $m_{\mrm{eff},c}$ are used to define the maximum allowed $z$ range, within which $z_a$ is chosen, together with the $m_a$ value. In the subsequent evolution of $b$ and $c$, only pairs of $m_b$ and $m_c$ are allowed for which the already selected $z_a$ fulfils the constraints in eq.~(\ref{sh:zbounds}). For `unconstrained' evolution, which is the preferred alternative, one may start off by assuming the daughters to be massless, so that the allowed $z$ range is \begin{equation} z_{\pm} = \frac{1}{2} \left\{ 1 \pm \frac{|\mbf{p}_a|}{E_a} \theta(m_a - m_{\mmin,a}) \right\} ~, \label{sh:zrange} \end{equation} where $\theta(x)$ is the step function, $\theta(x) = 1$ for $x > 0$ and $\theta(x) = 0$ for $x < 0$. The decay kinematics into two massless four-vectors $p_b^{(0)}$ and $p_c^{(0)}$ is now straightforward. Once $m_b$ and $m_c$ have been found from the subsequent evolution, subject only to the constraints $m_b < z_a E_a$, $m_c < (1-z_a) E_a$ and $m_b + m_c < m_a$, the actual massive four-vectors may be defined as \begin{equation} p_{b,c} = p_{b,c}^{(0)} \pm (r_c p_c^{(0)} - r_b p_b^{(0)}) ~, \label{sh:pshowershift} \end{equation} where \begin{equation} r_{b,c} = \frac{m_a^2 \pm (m_c^2 -m_b^2) - \sqrt{ (m_a^2 - m_b^2 - m_c^2)^2 - 4 m_b^2 m_c^2}}{2 m_a^2} ~. \end{equation} In other words, the meaning of $z_a$ is somewhat reinterpreted {\it post facto}. Needless to say, the `unconstrained' option allows more branchings to take place than the `constrained' one. In the following discussion we will only refer to the `global, unconstrained' $z$ choice. \subsubsection{First branchings and matrix-element matching} The final-state evolution is normally started from some initial parton pair $1 + 2$, at a $Q_{\mmax}^2$ scale determined by deliberations already discussed. When the evolution of parton 1 is considered, it is assumed that parton 2 is massless, so that the parton 1 energy and momentum are simple functions of its mass (and of the c.m. energy of the pair, which is fixed), and hence also the allowed $z_1$ range for splittings is a function of this mass, eq. (\ref{sh:zrange}). Correspondingly, parton 2 is evolved under the assumption that parton 1 is massless. After both partons have been assigned masses, their correct energies may be found, which are smaller than originally assumed. Therefore the allowed $z$ ranges have shrunk, and it may happen that a branching has been assigned a $z$ value outside this range. If so, the parton is evolved downwards in mass from the rejected mass value; if both $z$ values are rejected, the parton with largest mass is evolved further. It may also happen that the sum of $m_1$ and $m_2$ is larger than the c.m. energy, in which case the one with the larger mass is evolved downwards. The checking and evolution steps are iterated until an acceptable set of $m_1$, $m_2$, $z_1$ and $z_2$ has been found. The procedure is an extension of the veto algorithm, where an initial overestimation of the allowed $z$ range is compensated by rejection of some branchings. One should note, however, that the veto algorithm is not strictly applicable for the coupled evolution in two variables ($m_1$ and $m_2$), and that therefore some arbitrariness is involved. This is manifest in the choice of which parton will be evolved further if both $z$ values are unacceptable, or if the mass sum is too large. For quark and lepton pairs which come from the decay of a colour-singlet particle, the first branchings are matched to the explicit first-order matrix elements for gauge boson decays. This is also done, e.g. in $\H^0$ decays, which has spin 0 rather than 1, and for which in principle therefore the matrix elements are slightly different. The matching is based on a mapping of the parton-shower variables on to the 3-jet phase space. To produce a 3-jet event, $\gammaZ \to \q(p_1) \qbar(p_2) \g(p_3)$, in the shower language, one will pass through an intermediate state, where either the $\q$ or the $\qbar$ is off the mass shell. If the former is the case then \begin{eqnarray} m^2 & = & (p_1 + p_3)^2 = E_{\mrm{cm}}^2 (1 - x_2) ~, \nonumber \\ z & = & \frac{E_1}{E_1 + E_3} = \frac{x_1}{x_1 + x_3} = \frac{x_1}{2-x_2} ~, \label{sh:MEfrPS} \end{eqnarray} where $x_i = 2 E_i/E_{\mrm{cm}}$. The $\qbar$ emission case is obtained with $1 \leftrightarrow 2$. The parton-shower splitting expression in terms of $m^2$ and $z$, eq. (\ref{sh:Patobc}), can therefore be translated into the following differential 3-jet rate: \begin{eqnarray} \frac{1}{\sigma} \, \frac{\d \sigma_{\mrm{PS}}}{\d x_1 \, \d x_2} & = & \frac{\alphas}{2 \pi} \, C_F \, \frac{1}{(1-x_1)(1-x_2)} \times \nonumber \\ & \times & \left\{ \frac{1-x_1}{x_3} \left( 1 + \left( \frac{x_1}{2-x_2} \right)^2 \right) + \frac{1-x_2}{x_3} \left( 1 + \left( \frac{x_2}{2-x_1} \right)^2 \right) \right\} ~, \label{sh:PSwt} \end{eqnarray} where the first term inside the curly bracket comes from emission off the quark and the second term from emission off the antiquark. The corresponding expression in matrix-element language is \begin{equation} \frac{1}{\sigma} \, \frac{\d \sigma_{\mrm{ME}}}{\d x_1 \, \d x_2} = \frac{\alphas}{2 \pi} \, C_F \, \frac{1}{(1-x_1)(1-x_2)} \left\{ x_1^2 +x_2^2 \right\} ~. \label{sh:MEwt} \end{equation} With the kinematics choice of {\sc Jetset}, the matrix-element expression is always smaller than the parton-shower one. It is therefore possible to run the shower as usual, but to impose an extra weight factor $\d \sigma_{\mrm{ME}} / \d \sigma_{\mrm{PS}}$, which is just the ratio of the expressions in curly brackets. If a branching is rejected, the evolution is continued from the rejected $Q^2$ value onwards (the veto algorithm). The weighting procedure is applied to the first branching of both the $\q$ and the $\qbar$, in each case with the (nominal) assumption that none of the other partons branch (neither the sister nor the daughters), so that the relations of eq. (\ref{sh:MEfrPS}) are applicable. If a photon is emitted instead of a gluon, the emission rate in parton showers is given by \begin{eqnarray} \frac{1}{\sigma} \, \frac{\d \sigma_{\mrm{PS}}}{\d x_1 \, \d x_2} & = & \frac{\alphaem}{2 \pi} \, \frac{1}{(1-x_1)(1-x_2)} \times \nonumber \\ & \times & \left\{ e_{\q}^2 \, \frac{1-x_1}{x_3} \left( 1 + \left( \frac{x_1}{2-x_2} \right)^2 \right) + e_{\qbar}^2 \, \frac{1-x_2}{x_3} \left( 1 + \left( \frac{x_2}{2-x_1} \right)^2 \right) \right\} ~, \end{eqnarray} and in matrix elements by \cite{Gro81} \begin{equation} \frac{1}{\sigma} \, \frac{\d \sigma_{\mrm{ME}}}{\d x_1 \, \d x_2} = \frac{\alphaem}{2 \pi} \, \frac{1}{(1-x_1)(1-x_2)} \left\{ \left( e_{\q} \, \frac{1-x_1}{x_3} - e_{\qbar} \, \frac{1-x_2}{x_3} \right)^2 \left( x_1^2 +x_2^2 \right) \right\} ~. \end{equation} As in the gluon emission case, a weighting factor $\d \sigma_{\mrm{ME}} / \d \sigma_{\mrm{PS}}$ can therefore be applied when either the original $\q$ ($\ell$) or the original $\qbar$ ($\br{\ell}$) emits a photon. For a neutral resonance, such as $\Z^0$, where $e_{\qbar} = - e_{\q}$, the above expressions simplify and one recovers exactly the same ratio $\d \sigma_{\mrm{ME}} / \d \sigma_{\mrm{PS}}$ as for gluon emission. Compared with the standard matrix-element treatment, a few differences remain. The shower one automatically contains the Sudakov form factor and an $\alphas$ running as a function of the $\pT^2$ scale of the branching. The shower also allows all partons to evolve further, which means that the na\"{\i}ve kinematics assumed for a comparison with matrix elements is modified by subsequent branchings, e.g. that the energy of parton 1 is reduced when parton 2 is assigned a mass. All these effects are formally of higher order, and so do not affect a first-order comparison. This does not mean that the corrections need be small, but experimental results are encouraging: the approach outlined does every bit as good as explicit second-order matrix elements for the description of 4-jet production. \subsubsection{Subsequent branches and angular ordering} The shower evolution is (almost) always done on a pair of partons, so that energy and momentum can be conserved. In the first step of the evolution, the two original partons thus undergo branchings $1 \to 3 + 4$ and $2 \to 5 + 6$. As described above, the allowed $m_1$, $m_2$, $z_1$ and $z_2$ ranges are coupled by kinematical constraints. In the second step, the pair $3 + 4$ is evolved and, separately, the pair $5 + 6$. Considering only the former (the latter is trivially obtained by symmetry), the partons thus have nominal initial energies $E_3^{(0)} = z_1 E_1$ and $E_4^{(0)} = (1-z_1) E_1$, and maximum allowed virtualities $m_{\mmax,3} = \min(m_1,E_3^{(0)})$ and $m_{\mmax,4} = \min(m_1,E_4^{(0)})$. Initially partons 3 and 4 are evolved separately, giving masses $m_3$ and $m_4$ and splitting variables $z_3$ and $z_4$. If $m_3 + m_4 > m_1$, the parton of 3 and 4 that has the largest ratio of $m_i/m_{\mmax,i}$ is evolved further. Thereafter eq. (\ref{sh:pshowershift}) is used to construct corrected energies $E_3$ and $E_4$, and the $z$ values are checked for consistency. If a branching has to be rejected because the change of parton energy puts $z$ outside the allowed range, the parton is evolved further. This procedure can then be iterated for the evolution of the two daughters of parton 3 and for the two of parton 4, etc., until each parton reaches the cut-off mass $m_{\mmin}$. Then the parton is put on the mass shell. The model, as described so far, produces so-called conventional showers, wherein masses are strictly decreasing in the shower evolution. Emission angles are decreasing only in an average sense, however, which means that also fairly `late' branchings can give partons at large angles. Theoretical studies beyond the leading-log level show that this is not correct \cite{Mue81}, but that destructive interference effects are large in the region of non-ordered emission angles. To a very good first approximation, these so-called coherence effects can be taken into account in parton shower programs by requiring a strict ordering in terms of decreasing emission angles. The coherence phenomenon is known already from QED. One manifestation is the Chudakov effect \cite{Chu55}, discovered in the study of high-energy cosmic $\gamma$ rays impinging on a nuclear target. If a $\gamma$ is converted into a highly collinear $\ee$ pair inside the emulsion, the $\e^+$ and $\e^-$ in their travel through the emulsion ionize atoms and thereby produce blackening. However, near the conversion point the blackening is small: the $\e^+$ and $\e^-$ then are still close together, so that an atom traversed by the pair does not resolve the individual charges of the $\e^+$ and the $\e^-$, but only feels a net charge close to zero. Only later, when the $\e^+$ and $\e^-$ are separated by more than a typical atomic radius, are the two able to ionize independently of each other. The situation is similar in QCD, but is further extended, since now also gluons carry colour. For example, in a branching $\q_0 \to \q\g$ the $\q$ and $\g$ share a colour--anticolour pair, and therefore the $\q$ and $\g$ cannot emit subsequent gluons incoherently. Again the net effect is to reduce the amount of soft gluon emission: since a soft gluon (emitted at large angles) corresponds to a large (transverse) wavelength, the soft gluon is unable to resolve the separate colour charges of the $\q$ and the $\g$, and only feels the net charge carried by the $\q_0$. Such a soft gluon $\g'$ (in the region $\theta_{\q_0 \g'} > \theta_{\q \g}$) could therefore be thought of as being emitted by the $\q_0$ rather than by the $\q$--$\g$ system. If one considers only emission that should be associated with the $\q$ or the $\g$, to a good approximation, there is a complete destructive interference in the regions of non-decreasing opening angles, while partons radiate independently of each other inside the regions of decreasing opening angles ($\theta_{q g'} < \theta_{q g}$ and $\theta_{g g'} < \theta_{q g}$), once azimuthal angles are averaged over. The details of the colour interference pattern are reflected in non-uniform azimuthal emission probabilities. The first branchings of the shower are not affected by the angular-ordering requirement --- since the evolution is performed in the c.m. frame of the original parton pair, where the original opening angle is 180$^{\circ}$, any angle would anyway be smaller than this --- but here instead the matrix-element matching procedure is used, where applicable. Subsequently, each opening angle is compared with that of the preceding branching in the shower. For a branching $a \to bc$ the kinematical approximation \begin{equation} \theta_a \approx \frac{p_{\perp b}}{E_b} + \frac{p_{\perp c}}{E_c} \approx \sqrt{z_a (1-z_a)} m_a \left( \frac{1}{z_a E_a} + \frac{1}{(1-z_a) E_a} \right) = \frac{1}{\sqrt{z_a(1-z_a)}} \frac{m_a}{E_a} \end{equation} is used to derive the opening angle (this is anyway to the same level of approximation as the one in which angular ordering is derived). With $\theta_b$ of the $b$ branching calculated similarly, the requirement $\theta_b < \theta_a$ can be reduced to \begin{equation} \frac{z_b (1-z_b)}{m_b^2} > \frac{1-z_a}{z_a m_a^2} ~. \end{equation} Since photons do not obey angular ordering, the check on angular ordering is not performed when a photon is emitted. When a gluon is emitted in the branching after a photon, its emission angle is restricted by that of the preceding QCD branching in the shower, i.e. the photon emission angle does not enter. \subsubsection{Other final-state shower aspects} The electromagnetic coupling constant for the emission of photons on the mass shell is $\alphaem = \alphaem(Q^2 = 0) \approx 1/137$. For the strong coupling constant several alternatives are available, the default being the first-order expression $\alphas(\pT^2)$, where $\pT^2$ is defined by the approximate expression $\pT^2 \approx z(1-z) m^2$. Studies of next-to-leading-order corrections favour this choice \cite{Ama80}. The other alternatives are a fixed $\alphas$ and an $\alphas(m^2)$. With the default choice of $\pT^2$ as scale in $\alphas$, a further cut-off is introduced on the allowed phase space of gluon emission, not present in the options with fixed $\alphas$ or with $\alphas(m^2)$, nor in the QED shower. A minimum requirement, to ensure a well-defined $\alphas$, is that $\pT / \Lambda > 1.1$, but additionally {\Je} requires that $\pT > Q_0/2$. This latter requirement is not a necessity, but it makes sense when $\pT$ is taken to be the preferred scale of the branching process, rather than e.g. $m$. It reduces the allowed $z$ range, compared with the purely kinematical constraints. Since the $\pT$ cut is not present for photon emission, the relative ratio of photon to gluon emission off a quark is enhanced at small virtualities compared with na\"{\i}ve expectations; in actual fact this enhancement is largely compensated by the running of $\alphas$, which acts in the opposite direction. The main consequence, however, is that the gluon energy spectrum is peaked at around $Q_0$ and rapidly vanishes for energies below that, whilst the photon spectum extends all the way to zero energy. Previously it was said that azimuthal angles in branchings are chosen isotropically. In fact, as an option, it is possible to include some effects of gluon polarization, which correlate the production and the decay planes of a gluon, such that a $\g \to \g\g$ branching tends to take place in the production plane of the gluon, while a decay out of the plane is favoured for $\g \to \q\qbar$. The formulae are given e.g. in ref. \cite{Web86}, as simple functions of the $z$ value at the vertex where the gluon is produced and of the $z$ value when it branches. Also coherence phenomena lead to non-isotropic azimuthal distributions \cite{Web86}, which are included as a further option. In either case the $\varphi$ azimuthal variable is first chosen isotropically, then the weight factor due to polarization times coherence is evaluated, and the $\varphi$ value is accepted or rejected. In case of rejection, a new $\varphi$ is generated, and so on. While the rule is to have an initial pair of partons, there are a few examples where one or three partons have to be allowed to shower. If only one parton is given, it is not possible to conserve both energy and momentum. The choice has been made to conserve energy and jet direction, but the momentum vector is scaled down when the radiating parton acquires a mass. The `rest frame of the system', used e.g. in the $z$ definition, is taken to be whatever frame the jet is given in. In $\Upsilon \to \g\g\g$ decays and other primary three-parton configurations, one is left with the issue how the energy sharing variables $x_1$ and $x_2$ from the massless matrix elements should be reinterpreted for a massive three-parton configuration. We have made the arbitrary choice of preserving the energy of each parton, which means that relative angles between the original partons is changed. Mass triplets outside the allowed phase space are rejected and the evolution continued. Finally, it should be noted that two toy shower models are included as options. One is a scalar gluon model, in which the $\q \to \q\g$ branching kernel is replaced by $P_{\q \to \q\g}(z) = \frac{2}{3} (1-z)$. The couplings of the gluon, $\g \to \g\g$ and $\g \to \q\qbar$, have been left as free parameters, since they depend on the colour structure assumed in the model. The spectra are flat in $z$ for a spin 0 gluon. Higher-order couplings of the type $\g \to \g\g\g$ could well contribute significantly, but are not included. The second toy model is an Abelian vector one. In this option $\g \to \g \g$ branchings are absent, and $\g \to \q \qbar$ ones enhanced. More precisely, in the splitting kernels, eq. (\ref{sh:APker}), the Casimir factors are changed as follows: $C_F = 4/3 \to 1$, $N_C = 3 \to 0$, $T_R = n_f/2 \to 3n_f$. When using either of these options, one should be aware that also a number of other components in principle should be changed, from the running of $\alphas$ to the whole concept of fragmentation. One should therefore not take them too seriously. \subsection{Initial-State Showers} The initial-state showe algorithm in {\Py} is not quite as sophisticated as the final-state one. This is partly because initial-state radiation is less well understood theoretically, partly because the programming task is more complicated and ambiguous. Still, the program at disposal is known to do a reasonably good job of describing existing data, such as $\Z^0$ production properties at hadron colliders \cite{Sjo85}. \subsubsection{The shower structure} \label{sss:initshowstruc} A fast hadron may be viewed as a cloud of quasireal partons. Similarly a fast lepton may be viewed as surrounded by a cloud of photons and partons; in the program the two situations are on an equal footing, but here we choose the hadron as example. At each instant, an individual parton can initiate a virtual cascade, branching into a number of partons. This cascade can be described in terms of a tree-like structure, composed of many subsequent branchings $a \to bc$. Each branching involves some relative transverse momentum between the two daughters. In a language where four-momentum is conserved at each vertex, this implies that at least one of the $b$ and $c$ partons must have a space-like virtuality, $m^2 < 0$. Since the partons are not on the mass shell, the cascade only lives a finite time before reassembling, with those parts of the cascade that are most off the mass shell living the shortest time. A hard scattering, e.g. in deep inelastic leptoproduction, will probe the hadron at a given instant. The probe, i.e. the virtual photon in the leptoproduction case, is able to resolve fluctuations in the hadron up to the $Q^2$ scale of the hard scattering. Thus probes at different $Q^2$ values will seem to see different parton compositions in the hadron. The change in parton composition with $t = \ln(Q^2/\Lambda^2)$ is given by the evolution equations \begin{equation} \frac{\d f_b(x,t)}{\d t} = \sum_{a,c} \int \frac{\d x'}{x'} \, f_a(x',t) \, \frac{\alpha_{abc}}{2 \pi} \, P_{a \to bc} \left( \frac{x}{x'} \right) ~. \label{sh:sfevol} \end{equation} Here the $f_i(x,t)$ are the parton-distribution functions, expressing the probability of finding a parton $i$ carrying a fraction $x$ of the total momentum if the hadron is probed at virtuality $Q^2$. The $P_{a \to bc}(z)$ are given in eq. (\ref{sh:APker}). As before, $\alpha_{abc}$ is $\alphas$ for QCD shower and $\alphaem$ for QED ones. Eq. (\ref{sh:sfevol}) is closely related to eq. (\ref{sh:Patobc}): $\d {\cal P}_a$ describes the probability that a given parton $a$ will branch (into partons $b$ and $c$), $\d f_b$ the influx of partons $b$ from the branchings of partons $a$. (The expression $\d f_b$ in principle also should contain a loss term for partons $b$ that branch; this term is important for parton-distribution evolution, but does not appear explicitly in what we shall be using eq. (\ref{sh:sfevol}) for.) The absolute form of hadron parton distributions cannot be predicted in perturbative QCD, but rather have to be parametrized at some $Q_0$ scale, with the $Q^2$ dependence thereafter given by eq. (\ref{sh:sfevol}). Available parametrizations are discussed in section \ref{ss:structfun}. The lepton and photon parton distributions inside a lepton can be fully predicted, but here for simplicity are treated on equal footing with hadron parton distributions. If a hard interaction scatters a parton out of the incoming hadron, the `coherence' \cite{Gri83} of the cascade is broken: the partons can no longer reassemble completely back to the cascade-initiating parton. In this semiclassical picture, the partons on the `main chain' of consecutive branchings that lead directly from the initiating parton to the scattered parton can no longer reassemble, whereas fluctuations on the `side branches' to this chain may still disappear. A convenient description is obtained by assigning a space-like virtuality to the partons on the main chain, in such a way that the partons on the side branches may still be on the mass shell. Since the momentum transfer of the hard process can put the scattered parton on the mass shell (or even give it a time-like virtuality, so that it can initiate a final-state shower), one is then guaranteed that no partons have a space-like virtuality in the final state. (In real life, confinement effects obviously imply that partons need not be quite on the mass shell.) If no hard scattering had taken place, the virtuality of the space-like parton line would still force the complete cascade to reassemble. Since the virtuality of the cascade probed is carried by one single parton, it is possible to equate the space-like virtuality of this parton with the $Q^2$ scale of the cascade, to be used e.g. in the evolution equations. Further, coherence effects \cite{Gri83,Bas83} guarantee that the $Q^2$ vaules of the partons along the main chain are strictly ordered, with the largest $Q^2$ values close to the hard scattering. In recent years, further coherence effects have been studied \cite{Cia87}, with particular implications for the structure of parton showers at small $x$. None of these additional complications are implemented in the current algorithm, with the exception of a few rather primitive options that do not address the full complexity of the problem. Instead of having a tree-like structure, where all legs are treated democratically, the cascade is reduced to a single sequence of branchings $a \to bc$, where the $a$ and $b$ partons are on the main chain of space-like virtuality, $m_{a,b}^2 < 0$, while the $c$ partons are on the mass shell and do not branch. (Later we will include the possibility that the $c$ partons may have positive virtualities, $m_c^2 > 0$, which leads to the appearance of time-like `final-state' parton showers on the side branches.) This truncation of the cascade is only possible when it is known which parton actually partakes in the hard scattering: of all the possible cascades that exist virtually in the incoming hadron, the hard scattering will select one. To obtain the correct $Q^2$ evolution of parton distributions, e.g., it is essential that all branches of the cascade be treated democratically. In Monte Carlo simulation of space-like showers this is a major problem. If indeed the evolution of the complete cascade is to be followed from some small $Q_0^2$ up to the $Q^2$ scale of the hard scattering, it is no possible at the same time to handle kinematics exactly, since the virtuality of the various partons cannot be found until after the hard scattering has been selected. This kind of `forward evolution' scheme therefore requires a number of extra tricks to be made to work. Further, in this approach it is not known e.g. what the $\hat{s}$ of the hard scattering subsystem will be until the evolution has been carried out, which means that the initial-state evolution and the hard scattering have to be selected jointly, a not so trivial task. Instead we use the `backwards evolution' approach \cite{Sjo85}, in which the hard scattering is first selected, and the parton shower that preceded it is subsequently reconstructed. This reconstruction is started at the hard interaction, at the $Q_{\mmax}^2$ scale, and thereafter step by step one moves `backwards' in `time', towards smaller $Q^2$, all the way back to the parton-shower initiator at the cut-off scale $Q_0^2$. This procedure is possible if evolved parton distributions are used to select the hard scattering, since the $f_i(x,Q^2)$ contain the inclusive summation of all initial-state parton-shower histories that can lead to the appearance of an interacting parton $i$ at the hard scale. What remains is thus to select an exclusive history from the set of inclusive ones. \subsubsection{Longitudinal evolution} The evolution equations, eq. (\ref{sh:sfevol}), express that, during a small increase $\d t$ there is a probability for parton $a$ with momentum fraction $x'$ to become resolved into parton $b$ at $x = z x'$ and another parton $c$ at $x' - x = (1-z) x'$. Correspondingly, in backwards evolution, during a decrease $\d t$ a parton $b$ may be `unresolved' into parton $a$. The relative probability $\d {\cal P}_b$ for this to happen is given by the ratio $\d f_b / f_b$. Using eq. (\ref{sh:sfevol}) one obtains \begin{equation} \d {\cal P}_b = \frac{\d f_b(x,t)}{f_b(x,t)} = |\d t| \, \sum_{a,c} \int \frac{\d x'}{x'} \, \frac{f_a(x',t)}{f_b(x,t)} \, \frac{\alpha_{abc}}{2 \pi} \, P_{a \to bc} \left( \frac{x}{x'} \right) ~. \end{equation} Summing up the cumulative effect of many small changes $\d t$, the probability for no radiation exponentiates. Therefore one may define a form factor \begin{eqnarray} S_b(x,t_{\mmax},t) & = & \exp \left\{ - \int_t^{t_{\mmax}} \d t' \, \sum_{a,c} \int \frac{\d x'}{x'} \, \frac{f_a(x',t')}{f_b(x,t')} \, \frac{\alpha_{abc}(t')}{2\pi} \, P_{a \to bc} \left( \frac{x}{x'} \right) \right\} \nonumber \\ & = & \exp \left\{ - \int_t^{t_{\mmax}} \d t' \, \sum_{a,c} \int \d z \, \frac{\alpha_{abc}(t')}{2\pi} \, P_{a \to bc}(z) \, \frac{x'f_a(x',t')}{xf_b(x,t')} \right\} ~, \end{eqnarray} giving the probability that a parton $b$ remains at $x$ from $t_{\mmax}$ to a $t < t_{\mmax}$. It may be useful to compare this with the corresponding expression for forward evolution, i.e. with $S_a(t)$ in eq. (\ref{sh:Pnobranch}). The most obvious difference is the appearance of parton distributions in $S_b$. Structure functions are absent in $S_a$: the probability for a given parton $a$ to branch, once it exists, is independent of the density of partons $a$ or $b$. The parton distributions in $S_b$, on the other hand, express the fact that the probability for a parton $b$ to come from the branching of a parton $a$ is proportional to the number of partons $a$ there are in the hadron, and inversely proportional to the number of partons $b$. Thus the numerator $f_a$ in the exponential of $S_b$ ensures that the parton composition of the hadron is properly reflected. As an example, when a gluon is chosen at the hard scattering and evolved backwards, this gluon is more likely to have been emitted by a $\u$ than by a $\d$ if the incoming hadron is a proton. Similarly, if a heavy flavour is chosen at the hard scattering, the denominator $f_b$ will vanish at the $Q^2$ threshold of the heavy-flavour production, which means that the integrand diverges and $S_b$ itself vanishes, so that no heavy flavour remain below threshold. Another difference between $S_b$ and $S_a$, already touched upon, is that the $P_{\g \to \g\g}(z)$ splitting kernel appears with a normalization $2 N_C$ in $S_b$ but only with $N_C$ in $S_a$, since two gluons are produced but only one decays in a branching. A knowledge of $S_b$ is enough to reconstruct the parton shower backwards. At each branching $a \to bc$, three quantities have to be found: the $t$ value of the branching (which defines the space-like virtuality $Q_b^2$ of parton $b$), the parton flavour $a$ and the splitting variable $z$. This information may be extracted as follows: \begin{Enumerate} \item If parton $b$ partook in the hard scattering or branched into other partons at a scale $t_{\mmax}$, the probability that $b$ was produced in a branching $a \to bc$ at a lower scale $t$ is \begin{equation} \frac{\d {\cal P}_b}{\d t} = - \frac{\d S_b(x,t_{\mmax},t)}{\d t} = \left( \sum_{a,c} \int dz \, \frac{\alpha_{abc}(t')}{2\pi} \, P_{a \to bc}(z) \, \frac{x'f_a(x',t')}{xf_b(x,t')} \right) S_b(x,t_{\mmax},t) ~. \end{equation} If no branching is found above the cut-off scale $t_0$ the iteration is stopped and parton $b$ is assumed to be massless. \item Given the $t$ of a branching, the relative probabilities for the different allowed branchings $a \to bc$ are given by the $z$ integrals above, i.e. by \begin{equation} \int \d z \, \frac{\alpha_{abc}(t)}{2\pi} \, P_{a \to bc}(z) \, \frac{x'f_a(x',t)}{xf_b(x,t)} ~. \label{sh:Pbspace} \end{equation} \item Finally, with $t$ and $a$ known, the probability distribution in the splitting variable $z = x/x' = x_b/x_a$ is given by the integrand in eq. (\ref{sh:Pbspace}). \end{Enumerate} In addition, the azimuthal angle $\varphi$ of the branching is selected isotropically, i.e. no spin or coherence effects are included in this distribution. The selection of $t$, $a$ and $z$ is then a standard task of the kind than can be performed with the help of the veto algorithm. Specifically, upper and lower bounds for parton distributions are used to find simple functions that are everywhere larger than the integrands in eq. (\ref{sh:Pbspace}). Based on these simple expressions, the integration over $z$ may be carried out, and $t$, $a$ and $z$ values selected. This set is then accepted with a weight given by a ratio of the correct integrand in eq. (\ref{sh:Pbspace}) to the simple approximation used, both evaluated for the given set. Since parton distributions, as a rule, are not in a simple analytical form, it may be tricky to find reasonably good bounds to parton distributions. It is necessary to make different assumptions for valence and sea quarks, and be especially attentive close to a flavour threshold (\cite{Sjo85}). An electron distribution inside an electron behaves differently from parton distributions encountered in hadrons, and has to be considered separately. A comment on soft gluon emission. Nominally the range of the $z$ integral in $S_b$ is $x \leq z \leq 1$. The lower limit corresponds to $x' = x/z = 1$, and parton distributions vanish in this limit, wherefore no problems are encountered here. At the upper cut-off $z=1$ the splitting kernels $P_{\q \to \q\g}(z)$ and $P_{\g \to \g\g}$ diverge. This is the soft gluon singularity: the energy carried by the emitted gluon is vanishing, $x_{\g} = x' - x = (1-z) x' = (1-z) x/z \to 0$ for $z \to 1$. In order to calculate the integral over $z$ in $S_b$, an upper cut-off $z_{\mmax} = x/(x + x_{\epsilon})$ is introduced, i.e. only branchings with $z \leq z_{\mmax}$ are included in $S_b$. Here $x_{\epsilon}$ is a small number, typically chosen so that the gluon energy $x_{\g} \sqrt{s}/2 \geq x_{\epsilon} \sqrt{s}/2 = 2$ GeV. The average amount of energy carried away by gluons in the range $x_{g} < x_{\epsilon}$, over the given range of $t$ values from $t_a$ to $t_b$, may be estimated \cite{Sjo85}. The finally selected $z$ value may thus be picked as $z = z_{\mrm{hard}} \langle z_{\mrm{soft}}(t_a, t_b) \rangle$, where $z_{\mrm{hard}}$ is the originally selected $z$ value and $z_{\mrm{soft}}$ is the correction factor for soft gluon emission. In QED showers, the smallness of $\alphaem$ means that one can use rather smaller cut-off values without obtaining large amounts of emission. A fixed small cut-off $x_{\gamma} > 10^{-6}$ is therefore used to avoid the region of very soft photons. As has been discussed in section \ref{sss:estructfun}, the electron distribution inside the electron is cut off at $x_{\e} < 1 - 10^{-6}$, for numerical reasons, so the two cuts are closely matched. The cut-off scale $Q_0$ may be chosen separately for QCD and QED showers, just as in final-state radiation. The defaults are 1 GeV and 0.001 GeV, respectively. The former is the typical hadronic mass scale, below which radiation is not expected resolvable; the latter is of the order of the electron mass. Normally QED and QCD showers do not appear mixed. The most notable exception is resolved photoproduction (in $\ep$) and resolved 2$\gamma$ events (in $\ee$), i.e. shower histories of the type $\e \to \gamma \to \q$. Here the $Q^2$ scales need not be ordered at the interface, i.e. the last $\e \to \e\gamma$ branching may well have a larger $Q^2$ than the first $\q \to \q \g$ one, and the branching $\gamma \to \q$ does not even have a strict parton-shower interpretation for the vector dominance model part of the photon parton distribution. These issues are currently not addressed in full. Rather, based on the $x$ selected for the parton (quark or gluon) at the hard scattering, the $x_{\gamma}$ is selected once and for all in the range $x < x_{\gamma} <1$, according to the distribution implied by eq. (\ref{pg:foldqgine}). The QCD parton shower is then traced backwards from the hard scattering to the QCD shower initiator at $t_0$. No attempt is made to perform the full QED shower, but rather the beam remnant treatment (see section \ref{ss:beamrem}) is used to find the $\qbar$ (or $\g$) remnant that matches the $\q$ (or $\g$) QCD shower initiator, with the electron itself considered as a second beam remnant. \subsubsection{Transverse evolution} \label{sss:initshowtrans} We have above seen that two parton lines may be defined, stretching back from the hard scattering to the initial incoming hadron wavefunctions at small $Q^2$. Specifically, all parton flavours $i$, virtualities $Q^2$ and energy fractions $x$ may be found. The exact kinematical interpretation of the $x$ variable is not unique, however. For partons with small virtualities and transverse momenta, essentially all definitions agree, but differences may appear for branchings close to the hard scattering. In first-order QED \cite{Ber85} and in some simple QCD toy models \cite{Got86}, one may show that the `correct' choice is the `$\hat{s}$ approach'. Here one requires that $\hat{s} = x_1 x_2 s$, both at the hard scattering scale and at any lower scale, i.e. $\hat{s}(Q^2) = x_1(Q^2) \, x_2(Q^2) \, s$, where $x_1$ and $x_2$ are the $x$ values of the two resolved partons (one from each incoming beam particle) at the given $Q^2$ scale. In practice this means that, at a branching with the splitting variable $z$, the total $\hat{s}$ has to be increased by a factor $1/z$ in the backwards evolution. It also means that branchings on the two incoming legs have to be interleaved in a single monotonic sequence of $Q^2$ values of branchings. For a reconstruction of the complete kinematics in this approach, one should start with the hard scattering, for which $\hat{s}$ has been chosen according to the hard scattering matrix element. By backwards evolution, the virtualities $Q_1^2 = -m_1^2$ and $Q_2^2 = -m_2^2$ of the two interacting partons are reconstructed. Initially the two partons are considered in their common c.m. frame, coming in along the $\pm z$ directions. Then the four-momentum vectors have the non-vanishing components \begin{eqnarray} E_{1,2} & = & \frac{ \hat{s} \pm (Q_2^2 - Q_1^2)}{2 \sqrt{\hat{s}}} ~, \nonumber \\ p_{z1} = - p_{z2} & = & \sqrt{ \frac{ (\hat{s} + Q_1^2 + Q_2^2 )^2 - 4 Q_1^2 Q_2^2 }{ 4 \hat{s} } } ~, \end{eqnarray} with $(p_1 + p_2)^2 = \hat{s}$. If, say, $Q_1^2 > Q_2^2$, then the branching $3 \to 1 + 4$, which produced parton 1, is the one that took place closest to the hard scattering, and the one to be reconstructed first. With the four-momentum $p_3$ known, $p_4 = p_3 - p_1$ is automatically known, so there are four degrees of freedom. One corresponds to a trivial azimuthal angle around the $z$ axis. The $z$ splitting variable for the $3 \to 1 + 4$ vertex is found as the same time as $Q_1^2$, and provides the constraint $(p_3 + p_2)^2 = \hat{s}/z$. The virtuality $Q_3^2$ is given by backwards evolution of parton 3. One degree of freedom remains to be specified, and this is related to the possibility that parton 4 initiates a time-like parton shower, i.e. may have a non-zero mass. The maximum allowed squared mass $m_{\mmax,4}^2$ is found for a collinear branching $3 \to 1 + 4$. In terms of the combinations \begin{eqnarray} s_1 & = & \hat{s} + Q_2^2 + Q_1^2 ~, \nonumber \\ s_3 & = & \frac{\hat{s}}{z} + Q_2^2 + Q_3^2 ~, \nonumber \\ r_1 & = & \sqrt{s_1^2 - 4 Q_2^2 Q_1^2} ~, \nonumber \\ r_3 & = & \sqrt{s_3^2 - 4 Q_2^2 Q_3^2} ~, \end{eqnarray} one obtains \begin{equation} m_{\mmax,4}^2 = \frac{s_1 s_3 - r_1 r_3}{2 Q_2^2} - Q_1^2 - Q_3^2 ~, \end{equation} which, for the special case of $Q_2^2 = 0$, reduces to \begin{equation} m_{\mmax,4}^2 = \left\{ \frac{Q_1^2}{z} - Q_3^2 \right\} \left\{ \frac{\hat{s}}{\hat{s} + Q_1^2} - \frac{\hat{s}}{\hat{s}/z + Q_3^2} \right\} ~. \label{sh:zrangespace} \end{equation} These constraints on $m_4$ are only the kinematical ones, in addition coherence phenomena could constrain the $m_{\mmax,4}$ values further. Some options of this kind are available; the default one is to require additionally that $m_4^2 \leq Q_1^2$, i.e. lesser than the space-like virtuality of the sister parton. With the maximum virtuality given, the final-state showering machinery may be used to give the development of the subsequent cascade, including the actual mass $m_4^2$, with $0 \leq m_4^2 \leq m_{\mmax,4}^2$. The evolution is performed in the c.m. frame of the two `resolved' partons, i.e. that of partons 1 and 2 for the branching $3 \to 1 + 4$, and parton 4 is assumed to have a nominal energy $E_{\mrm{nom},4} = (1/z - 1) \sqrt{\hat{s}}/2$. (Slight modifications appear if parton 4 has a non-vanishing mass $m_{\q}$ or $m_{\ell}$.) Using the relation $m_4^2 = (p_3 - p_1)^2$, the momentum of parton 3 may now be found as \begin{eqnarray} E_3 & = & \frac{1}{2 \sqrt{\hat{s}} } \left\{ \frac{\hat{s}}{z} + Q_2^2 - Q_1^2 - m_4^2 \right\} ~, \nonumber \\ p_{z3} & = & \frac{1}{2 p_{z1}} \left\{ s_3 - 2 E_2 E_3 \right\} ~, \nonumber \\ p_{\perp ,3}^2 & = & \left\{ m_{\mmax,4}^2 - m_4^2 \right\} \, \frac{ (s_1 s_3 + r_1 r_3)/2 - Q_2^2 (Q_1^2 + Q_3^2 + m_4^2)}{r_1^2} ~. \end{eqnarray} The requirement that $m_4^2 \geq 0$ (or $\geq m_f^2$ for heavy flavours) imposes a constraint on allowed $z$ values. This constraint cannot be included in the choice of $Q_1^2$, where it logically belongs, since it also depends on $Q_2^2$ and $Q_3^2$, which are unknown at this point. It is fairly rare (in the order of 10\% of all events) that an unallowed $z$ value is generated, and when it happens it is almost always for one of the two branchings closest to the hard interaction: for $Q_2^2 = 0$ eq. (\ref{sh:zrangespace}) may be solved to yield $z \leq \hat{s}/(\hat{s} + Q_1^2 - Q_3^2)$, which is a more severe cut for $\hat{s}$ small and $Q_1^2$ large. Therefore an essentially bias-free way of coping is to redo completely any initial-state cascade for which this problem appears. This completes the reconstruction of the $3 \to 1 + 4$ vertex. The subsystem made out of partons 3 and 2 may now be boosted to its rest frame and rotated to bring partons 3 and 2 along the $\pm z$ directions. The partons 1 and 4 now have opposite and compensating transverse momenta with respect to the event axis. When the next vertex is considered, either the one that produces parton 3 or the one that produces parton 2, the 3--2 subsystem will fill the function the 1--2 system did above, e.g. the r\^ole of $\hat{s} = \hat{s}_{12}$ in the formulae above is now played by $\hat{s}_{32} = \hat{s}_{12}/z$. The internal structure of the 3--2 system, i.e. the branching $3 \to 1 + 4$, appears nowhere in the continued description, but has become `unresolved'. It is only reflected in the successive rotations and boosts performed to bring back the new endpoints to their common rest frame. Thereby the hard scattering subsystem 1--2 builds up a net transverse momentum and also an overall rotation of the hard scattering subsystem. After a number of steps, the two outermost partons have virtualities $Q^2 < Q_0^2$ and then the shower is terminated and the endpoints assigned $Q^2 = 0$. Up to small corrections from primordial $k_{\perp}$, discussed in section \ref{ss:beamrem}, a final boost will bring the partons from their c.m. frame to the overall c.m. frame, where the $x$ values of the outermost partons agree also with the light-cone definition. \subsubsection{Other initial-state shower aspects} In the formulae above, $Q^2$ has been used as argument for $\alphas$, and not only as the space-like virtuality of partons. This is one possibility, but in fact loop calculations tend to indicate that the proper argument for $\alphas$ is not $Q^2$ but $\pT^2 = (1-z) Q^2$ \cite{Bas83}. The variable $\pT$ does have the interpretation of transverse momentum, although it is only exactly so for a branching $a \to bc$ with $a$ and $c$ massless and $Q^2 = - m_b^2$, and with $z$ interpreted as light-cone fraction of energy and momentum. The use of $\alphas((1-z)Q^2)$ is default in the program. Indeed, if one wanted to, the complete shower might be interpreted as an evolution in $\pT^2$ rather than in $Q^2$. As we see, the initial-state showering algorithm leads to a net boost and rotation of the hard scattering subsystems. The overall final state is made even more complex by the additional final-state radiation. In principle, the complexity is very physical, but it may still have undesirable side effects. One such, discussed further in section \ref{ss:PYswitchkin}, is that it is very difficult to generate events that fulfill specific kinematics conditions, since kinematics is smeared and even, at times, ambiguous. A special case is encountered in deep inelastic scattering in $\ep$ collisions. Here the DIS $x$ and $Q^2$ values are defined in terms of the scattered electron direction and energy, and therefore are unambiguous (except for issues of final-state photon radiation close to the electron direction). Neither initial- nor final-state showers preserve the kinematics of the scattered electron, however, and hence the DIS $x$ and $Q^2$ are changed. In principle, this is perfectly legitimate, with the caveat that one then also should use different sets of parton distributions than ones derived from DIS, since these are based on the kinematics of the scattered lepton and nothing else. Alternatively, one might consider showering schemes that leave $x$ and $Q^2$ unchanged. In \cite{Ben88} detailed modifications are presented that make a preservation possible when radiation off the incoming and outgoing electron is neglected, but these are not included in the current version of {\Py}. What is available, as an option, is a simple machinery which preserves $x$ and $Q^2$ from the effects of QCD radiation, and also from those of primordial $k_{\perp}$ and the beam remnant treatment, as follows. After the showers have been generated, the four-momentum of the scattered lepton is changed to the expected one, based on the nominal $x$ and $Q^2$ values. The azimuthal angle of the lepton is maintained when the transverse momentum is adjusted. Photon radiation off the lepton leg is not fully accounted for, i.e. it is assumed that the energy of final-state photons is added to that of the scattered electron for the definition of $x$ and $Q^2$ (this is the normal procedure for parton-distribution definitions). The change of three-momentum on the lepton side of the event is balanced by the final state partons on the hadron side, excluding the beam remnant but including all the partons both from initial- and final-state showering. The fraction of three-momentum shift taken by each parton is proportional to its original light-cone momentum in the direction of the incoming lepton, i.e. to $E \mp p_z$ for a hadron moving in the $\pm$ direction. This procedure guarantees momentum but not energy conservation. For the latter, one additional degree of freedom is needed, which is taken to be the longitudinal momentum of the initial state shower initiator. As this momentum is modified, the change is shared by the final state partons on the hadron side, according to the same light-cone fractions as before (based on the original momenta). Energy conservation requires that the total change in final state parton energies plus the change in lepton side energy equals the change in initiator energy. This condition can be turned into an iterative procedure to find the initiator momentum shift. Sometimes the procedure may break down. For instance, an initiator with $x > 1$ may be reconstructed. If this should happen, the $x$ and $Q^2$ values of the event are preserved, but new initial and final state showers are generated. After five such failures, the event is completely discared in favour of a new kinematical setup. Kindly note that the four-momentum of intermediate partons in the shower history are not being adjusted. In a listing of the complete event history, energy and momentum need then not be conserved in shower branchings. This mismatch could be fixed up, if need be. The scheme presented above should not be taken too literally, but is rather intended as a contrast to the more sophisticated schemes already on the market, if one would like to understand whether the kind of conservation scheme chosen does affect the observable physics. \subsection{Routines and Common Block Variables} \label{ss:showrout} In this section we collect information on how to use the initial- and final-state showering routines. Of these {\tt LUSHOW} for final-state radiation is the more generally interesting, since it can be called to let a user-defined parton configuration shower. {\tt PYSSPA}, on the other hand, is so intertwined with the general structure of a {\Py} event that it is of little use as a stand-alone product. \drawbox{CALL LUSHOW(IP1,IP2,QMAX)}\label{p:LUSHOW} \begin{entry} \itemc{Purpose:} to generate time-like parton showers, conventional or coherent. The performance of the program is regulated by the switches {\tt MSTJ(41) - MSTJ(50)} and parameters {\tt PARJ(82) - PARJ(84)}. In order to keep track of the colour flow information, the positions {\tt K(I,4)} and {\tt K(I,5)} have to be organized properly for showering partons. Inside the {\JePy} programs, this is done automatically, but for external use proper care must be taken. \iteme{IP1 > 0, IP2 = 0 :} generate a time-like parton shower for the parton in line {\tt IP1} in common block {\tt LUJETS}, with maximum allowed mass {\tt QMAX}. With only one parton at hand, one cannot simultaneously conserve both energy and momentum: we here choose to conserve energy and jet direction, while longitudinal momentum (along the jet axis) is not conserved. \iteme{IP1 > 0, IP2 > 0 :} generate time-like parton showers for the two partons in lines {\tt IP1} and {\tt IP2} in the common block {\tt LUJETS}, with maximum allowed mass for each parton {\tt QMAX}. For shower evolution, the two partons are boosted to their c.m. frame. Energy and momentum is conserved for the pair of partons, although not for each individually. One of the two partons may be replaced by a nonradiating particle, such as a photon or a diquark; the energy and momentum of this particle will then be modified to conserve the total energy and momentum. \iteme{IP1 > 0, IP2 < 0 :} generate time-like parton showers for the {\tt -IP2} (at most 3) partons in lines {\tt IP1}, {\tt IP1+1}, \ldots {\tt IPI-IP2-1} in the common block {\tt LUJETS}, with maximum allowed mass for each parton {\tt QMAX}. The actions for {\tt IP2=-1} and {\tt IP2=-2} correspond to what is described above, but additionally {\tt IP2=-3} may be used to generate the evolution starting from three given partons (e.g. in $\Upsilon \to \g\g\g$). Then the three partons are boosted to their c.m. frame, energy is conserved for each parton individually and momentum for the system as a whole. \iteme{QMAX :} the maximum allowed mass of a radiating parton, i.e. the starting value for the subsequent evolution. (In addition, the mass of a single parton may not exceed its energy, the mass of a parton in a system may not exceed the invariant mass of the system.) \end{entry} \boxsep \begin{entry} \iteme{SUBROUTINE PYSSPA(IPU1,IPU2) :}\label{p:PYSSPA} to generate the space-like showers of the initial-state radiation. \end{entry} \drawbox{COMMON/LUDAT1/MSTU(200),PARU(200),MSTJ(200),PARJ(200)} \begin{entry} \itemc{Purpose:} to give access to a number of status codes and parameters which regulate the performance of {\Je}. Most parameters are described in section \ref{ss:JETswitch}; here only those related to {\tt LUSHOW} are described. \boxsep \iteme{MSTJ(41) :}\label{p:MSTJ41} (D=2) type of branchings allowed in shower. \begin{subentry} \iteme{= 0 :} no branchings at all, i.e. shower is switched off. \iteme{= 1 :} QCD type branchings of quarks and gluons. \iteme{= 2 :} also emission of photons off quarks and leptons; the photons are assumed on the mass shell. \iteme{= 10 :} as {\tt =2}, but enhance photon emission by a factor {\tt PARJ(84)}. This option is unphysical, but for moderate values, {\tt PARJ(84)}$\leq 10$, it may be used to enhance the prompt photon signal in $\q\qbar$ events. The normalization of the prompt photon rate should then be scaled down by the same factor. The dangers of an improper use are significant, so do not use this option if you do not know what you are doing. \end{subentry} \iteme{MSTJ(42) :} (D=2) branching mode for time-like showers. \begin{subentry} \iteme{= 1 :} conventional branching, i.e. without angular ordering. \iteme{= 2 :} coherent branching, i.e. with angular ordering. \end{subentry} \iteme{MSTJ(43) :} (D=4) choice of $z$ definition in branching. \begin{subentry} \iteme{= 1 :} energy fraction in grandmother's rest frame (`local, constrained'). \iteme{= 2 :} energy fraction in grandmother's rest frame assuming massless daughters, with energy and momentum reshuffled for massive ones (`local, unconstrained'). \iteme{= 3 :} energy fraction in c.m. frame of the showering partons (`global, constrained'). \iteme{= 4 :} energy fraction in c.m. frame of the showering partons assuming massless daughters, with energy and momentum reshuffled for massive ones (`global, unconstrained'). \end{subentry} \iteme{MSTJ(44) :} (D=2) choice of $\alphas$ scale for shower. \begin{subentry} \iteme{= 0 :} fixed at {\tt PARU(111)} value. \iteme{= 1 :} running with $Q^2 = m^2/4$, $m$ mass of decaying parton, $\Lambda$ as stored in {\tt PARJ(81)} (natural choice for conventional showers). \iteme{= 2 :} running with $Q^2 = z(1-z)m^2$, i.e. roughly $\pT^2$ of branching, $\Lambda$ as stored in {\tt PARJ(81)} (natural choice for coherent showers). \end{subentry} \iteme{MSTJ(45) :} (D=5) maximum flavour that can be produced in shower by $\g \to \q \qbar$; also used to determine the maximum number of active flavours in the $\alphas$ factor in parton showers (here with a minimum of 3). \iteme{MSTJ(46) :} (D=3) nonhomogeneous azimuthal distributions in a shower branching. \begin{subentry} \iteme{= 0 :} azimuthal angle is chosen uniformly. \iteme{= 1 :} nonhomogeneous azimuthal angle in gluon decays due to a kinematics-dependent effective gluon polarization. Not meaningful for scalar model, i.e. then same as {\tt =0}. \iteme{= 2 :} nonhomogeneous azimuthal angle in gluon decay due to interference with nearest neighbour (in colour). Not meaningful for Abelian model, i.e. then same as {\tt =0}. \iteme{= 3 :} nonhomogeneous azimuthal angle in gluon decay due to both polarization ({\tt =1}) and interference ({\tt =2}). Not meaningful for Abelian model, i.e. then same as {\tt =1}. Not meaningful for scalar model, i.e. then same as {\tt =2}. \end{subentry} \iteme{MSTJ(47) :} (D=3) corrections to the lowest-order $\q\qbar\g$, $\q\qbar\gamma$, $\ell^+\ell^-\gamma$ or $\ell\nu_{\ell}\gamma$ 3-parton matrix element at the first branching of either initial parton in a shower. \begin{subentry} \iteme{= 0 :} no corrections. \iteme{= 1 :} included whenever scattered partons are $\q\qbar$, $\ell^+\ell^-$ or $\ell\nu_{\ell}$. \iteme{= 2 :} always included when shower starts from two partons. \iteme{= 3 :} as {\tt =1} except that for massive quarks also the massive matrix element expression is used, eq. (\ref{ee:threejMEmass}), while {\tt =1} is always based on massless matrix elements. \iteme{= 4 :} as {\tt =2} except that for massive quarks also the massive matrix element expression is used, while {\tt =2} is always based on massless matrix elements. \end{subentry} \iteme{MSTJ(48) :} (D=0) possibility to impose maximum angle for the first branching in a shower. \begin{subentry} \iteme{= 0 :} no explicit maximum angle. \iteme{= 1 :} maximum angle given by {\tt PARJ(85)} for single showering parton, by {\tt PARJ(85)} and {\tt PARJ(86)} for pair of showering partons. \end{subentry} \iteme{MSTJ(49) :} (D=0) possibility to change the branching probabilities according to some alternative toy models (note that the $Q^2$ evolution of $\alphas$ may well be different in these models, but that only the {\tt MSTJ(44)} options are at the disposal of the user). \begin{subentry} \iteme{= 0 :} standard QCD branchings. \iteme{= 1 :} branchings according to a scalar gluon theory, i.e. the splitting kernels in the evolution equations are, with a common factor $\alphas/(2\pi)$ omitted, $P_{\q \to \q\g} = (2/3) (1-z)$, $P_{\g \to \g\g} =$ {\tt PARJ(87)}, $P_{\g \to \q\qbar} =$ {\tt PARJ(88)} (for each separate flavour). The couplings of the gluon have been left as free parameters, since they depend on the colour structure assumed. Note that, since a spin 0 object decays isotropically, the gluon splitting kernels contain no $z$ dependence. \iteme{= 2 :} branchings according to an Abelian vector gluon theory, i.e. the colour factors are changed (compared with QCD) according to $C_F = 4/3 \to 1$, $N_C = 3 \to 0$, $T_R = 1/2 \to 3$. Note that an Abelian model is not expected to contain any coherence effects between gluons, so that one should normally use {\tt MSTJ(42)=1} and {\tt MSTJ(46)=} 0 or 1. Also, $\alphas$ is expected to increase with increasing $Q^2$ scale, rather than decrease. No such $\alphas$ option is available; the one that comes closest is {\tt MSTJ(44)=0}, i.e. a fix value. \end{subentry} \iteme{MSTJ(50) :} (D=3) possibility to introduce colour coherence effects in the first branching of a final state shower; mainly of relevance for QCD parton--parton scattering processes. \begin{subentry} \iteme{= 0 :} none. \iteme{= 1 :} impose an azimuthal anisotropy. \iteme{= 2 :} restrict the polar ange of a branching to be smaller than the scattering angle of the relevant colour flow. \iteme{= 3 :} both azimuthal anisotropy and restricted polar angles. \itemc{Note:} for subsequent branchings the (polar) angular ordering is automatic ({\tt MSTP(42)=2}) and {\tt MSTJ(46)=3}). \end{subentry} \boxsep \iteme{PARJ(81) :}\label{p:PARJ81} (D=0.29 GeV) $\Lambda$ value used in running $\alphas$ for parton showers (see {\tt MSTJ(44)}). \iteme{PARJ(82) :} (D=1.0 GeV) invariant mass cut-off $m_{\mmin}$ of parton showers, below which partons are not assumed to radiate. For $Q^2 = \pT^2$ ({\tt MSTJ(44)=2}) {\tt PARJ(82)}/2 additionally gives the minimum $\pT$ of a branching. To avoid infinite $\alphas$ values, one must have {\tt PARJ(82)}$ > 2 \times${\tt PARJ(81)} for {\tt MSTJ(44)}$\geq 1$ (this is automatically checked in the program, with $2.2 \times${\tt PARJ(81)} as the lowest value attainable). \iteme{PARJ(83) :} (D=1.0 GeV) invariant mass cut-off $m_{\mmin}$ used for photon emission in parton showers, below which quarks and leptons are not assumed to radiate. The function of {\tt PARJ(83)} closely parallels that of {\tt PARJ(82)} for QCD branchings, but there is a priori no requirement that the two be equal. \iteme{PARJ(84) :} (D=1.) used for option {\tt MSTJ(41)=10} as a multiplicative factor in the promt photon emission rate in final state parton showers. Unphysical but useful technical trick, so beware! \iteme{PARJ(85), PARJ(86) :} (D=10.,10.) maximum opening angles allowed in the first branching of parton showers; see {\tt MSTJ(48)}. \iteme{PARJ(87) :} (D=0.) coupling of $\g \to \g\g$ in scalar gluon shower, see {\tt MSTJ(49)=1}. \iteme{PARJ(88) :} (D=0.) coupling of $\g \to \q\qbar$ in scalar gluon shower (per quark species), see {\tt MSTJ(49)=1}. \end{entry} \drawbox{COMMON/PYPARS/MSTP(200),PARP(200),MSTI(200),PARI(200)} \begin{entry} \itemc{Purpose:} to give access to status code and parameters which regulate the performance of {\Py}. Most parameters are described in section \ref{ss:PYswitchpar}; here only those related to {\tt PYSSPA} and {\tt LUSHOW} are described. \iteme{MSTP(22) :}\label{p:MSTP22} (D=0) special override of normal $Q^2$ definition used for maximum of parton-shower evolution. This option only affects processes 10 and 83 (deep inelastic scattering) and only in lepton--hadron events. \begin{subentry} \iteme{= 0 :} use the scale as given in {\tt MSTP(32)}. \iteme{= 1 :} use the DIS $Q^2$ scale, i.e. $-\hat{t}$. \iteme{= 2 :} use the DIS $W^2$ scale, i.e. $(-\hat{t})(1-x)/x$. \iteme{= 3 :} use the DIS $Q \times W$ scale, i.e. $(-\hat{t}) \sqrt{(1-x)/x}$. \iteme{= 4 :} use the scale $Q^2 (1-x) \max(1, \ln(1/x))$, as motivated by first order matrix elements \cite{Ing80,Alt78}. \itemc{Note:} in all of these alternatives, a multiplicative factor is introduced by {\tt PARP(67)} and {\tt PARP(71)}, as usual. \end{subentry} \iteme{MSTP(61) :}\label{p:MSTP61} (D=1) master switch for initial-state QCD and QED radiation. \begin{subentry} \iteme{= 0 :} off. \iteme{= 1 :} on. \end{subentry} \iteme{MSTP(62) :} (D=3) level of coherence imposed on the space-like parton-shower evolution. \begin{subentry} \iteme{= 1 :} none, i.e. neither $Q^2$ values nor angles need be ordered. \iteme{= 2 :} $Q^2$ values at branches are strictly ordered, increasing towards the hard interaction. \iteme{= 3 :} $Q^2$ values and opening angles of emitted (on-mass-shell or time-like) partons are both strictly ordered, increasing towards the hard interaction. \end{subentry} \iteme{MSTP(63) :} (D=2) structure of associated time-like showers, i.e. showers initiated by emission off the incoming space-like partons. \begin{subentry} \iteme{= 0 :} no associated showers are allowed, i.e. emitted partons are put on the mass shell. \iteme{= 1 :} a shower may evolve, with maximum allowed time-like virtuality set by the phase space only. \iteme{= 2 :} a shower may evolve, with maximum allowed time-like virtuality set by phase space or by {\tt PARP(71)} times the $Q^2$ value of the space-like parton created in the same vertex, whichever is the stronger constraint. \end{subentry} \iteme{MSTP(64) :} (D=2) choice of $\alphas$ and $Q^2$ scale in space-like parton showers. \begin{subentry} \iteme{= 0 :} $\alphas$ is taken to be fix at the value {\tt PARU(111)}. \iteme{= 1 :} first-order running $\alphas$ with argument {\tt PARP(63)}$Q^2$. \iteme{= 2 :} first-order running $\alphas$ with argument {\tt PARP(64)}$k_{\perp}^2 = ${\tt PARP(64)}$(1-z)Q^2$. \end{subentry} \iteme{MSTP(65) :} (D=1) treatment of soft gluon emission in space-like parton-shower evolution. \begin{subentry} \iteme{= 0 :} soft gluons are entirely neglected. \iteme{= 1 :} soft gluon emission is resummed and included together with the hard radiation as an effective $z$ shift. \end{subentry} \iteme{MSTP(66) :} (D=1) choice of lower cut-off for initial-state QCD radiation in anomalous photoproduction events (see {\tt MSTP(14)=3}). \begin{subentry} \iteme{= 0 :} the lower $Q^2$ cutoff is the standard one in {\tt PARP(62)}$^2$. \iteme{= 1 :} the lower cutoff is the larger of {\tt PARP(62)}$^2$ and {\tt VINT(282)}, where the latter is virtuality scale of the $\gamma \to \q\qbar$ vertex. \end{subentry} \iteme{MSTP(67) :} (D=2) possibility to introduce colour coherence effects in the first branching of the backwards evolution of an initial state shower; mainly of relevance for QCD parton--parton scattering processes. \begin{subentry} \iteme{= 0 :} none. \iteme{= 2 :} restrict the polar angle of a branching to be smaller than the scattering angle of the relevant colour flow. \itemc{Note 1:} azimuthal anisotropies have not yet been included. \itemc{Note 2:} for subsequent branchings, {\tt MSTP(62)=3} is used to restrict the (polar) angular range of branchings. \end{subentry} \iteme{MSTP(71) :} (D=1) master switch for final-state QCD and QED radiation. \begin{subentry} \iteme{= 0 :} off. \iteme{= 1 :} on. \end{subentry} \boxsep \iteme{PARP(61) :}\label{p:PARP61} (D=0.25 GeV) $\Lambda$ value used in space-like parton shower (see {\tt MSTP(64)}). This value may be overwritten, see {\tt MSTP(3)}. \iteme{PARP(62) :} (D=1. GeV) effective cut-off $Q$ or $k_{\perp}$ value (see {\tt MSTP(64)}), below which space-like parton showers are not evolved. \iteme{PARP(63) :} (D=0.25) in space-like shower evolution the virtuality $Q^2$ of a parton is multiplied by {\tt PARP(63)} for use as a scale in $\alphas$ and parton distributions when {\tt MSTP(64)=1}. \iteme{PARP(64) :} (D=1.) in space-like parton-shower evolution the squared transverse momentum evolution scale $k_{\perp}^2$ is multiplied by {\tt PARP(64)} for use as a scale in $\alphas$ and parton distributions when {\tt MSTP(64)=2}. \iteme{PARP(65) :} (D=2. GeV) effective minimum energy (in c.m. frame) of time-like or on-shell parton emitted in space-like shower; see also {\tt PARP(66)}. \iteme{PARP(66) :} (D=0.001) effective lower cut-off on $1-z$ in space-like showers, in addition to the cut implied by {\tt PARP(65)}. \iteme{PARP(67) :} (D=4.) the $Q^2$ scale of the hard scattering (see {\tt MSTP(32)}) is multiplied by {\tt PARP(67)} to define the maximum parton virtuality allowed in space-like showers. This does not apply to $s$-channel resonances, where the maximum virtuality is set by $m^2$. \iteme{PARP(68) :} (D=1E-3) lower $Q$ cut-off for QED space-like showers. \iteme{PARP(71) :} (D=4.) the $Q^2$ scale of the hard scattering (see {\tt MSTP(32)}) is multiplied by {\tt PARP(71)} to define the maximum parton virtuality allowed in time-like showers. This does not apply to $s$-channel resonances, where the maximum virtuality is set by $m^2$. \end{entry} \clearpage \section{Beam Remnants and Underlying Events} Each incoming beam particle may leave behind a beam remnant, which does not take part in the initial-state radiation or hard scattering process. If nothing else, the remnants need be reconstructed and connected to the rest of the event. In hadron--hadron collisions, the composite nature of the two incoming beam particles implies the additional possibility that several parton pairs undergo separate hard or semi-hard scatterings, `multiple interactions'. This may give a non-negligible contribution to the `underlying event' structure, and thus to the total multiplicity. Finally, in high-luminosity colliders, it is possible to have several collisions between beam particles in one and the same beam crossing, i.e. pile-up events, which further act to build up the general particle production activity that is to be observed by detectors. These three aspects are described in turn, with emphasis on the middle one, that of multiple interactions within a single hadron--hadron collision. The main reference on the multiple interactions model is \cite{Sjo87a}. \subsection{Beam Remnants} \label{ss:beamrem} The initial-state radiation algorithm reconstructs one shower initiator in each beam. (If initial-state radiation is not included, the initiator is nothing but the incoming parton to the hard interaction.) Together the two initiators delineate an interaction subsystem, which contains all the partons that participate in the initial-state showers, in the hard interaction, and in the final-state showers. Left behind are two beam remnants which, to first approximation, just sail through, unaffected by the hard process. (The issue of additional interactions is covered in the next section.) A description of the beam remnant structure contains a few components. First, given the flavour content of a (colour-singlet) beam particle, and the flavour and colour of the initiator parton, it is possible to reconstruct the flavour and colour of the beam remnant. Sometimes the remnant may be represented by just a single parton or diquark, but often the remnant has to be subdivided into two separate objects. In the latter case it is necessary to share the remnant energy and momentum between the two. Due to Fermi motion inside hadron beams, the initiator parton may have a `primordial $k_{\perp}$' transverse momentum motion, which has to be compensated by the beam remnant. If the remnant is subdivided, there may also be a relative transverse momentum. In the end, total energy and momentum has to be conserved. To first approximation, this is ensured within each remnant separately, but some final global adjustments are necessary to compensate for the primordial $k_{\perp}$ and any effective beam remnant mass. Consider first a proton (or, with trivial modifications, any other baryon or antibaryon). \begin{Itemize} \item If the initiator parton is a $\u$ or $\d$ quark, it is assumed to be a valence quark, and therefore leaves behind a diquark beam remnant, i.e. either a $\u\d$ or a $\u\u$ diquark, in a colour antitriplet state. Relative probabilities for different diquark spins are derived within the context of the non-relativistic {\bf SU(6)} model, i.e. flavour {\bf SU(3)} times spin {\bf SU(2)}. Thus a $\u\d$ is $3/4$ $\u\d_0$ and $1/4$ $\u\d_1$, while a $\u\u$ is always $\u\u_1$. \item An initiator gluon leaves behind a colour octet $\u\u\d$ state, which is subdivided into a colour triplet quark and a colour antitriplet diquark. {\bf SU(6)} gives the appropriate subdivision, $1/2$ of the time into $\u + \u\d_0$, $1/6$ into $\u + \u\d_1$ and $1/3$ into $\d + \u\u_1$. \item A sea quark initiator, such as an $\s$, leaves behind a $\u\u\d\sbar$ four-quark state. The PDG flavour coding scheme and the fragmentation routines do not foresee such a state, so therefore it is subdivided into a meson plus a diquark, i.e. $1/2$ into $\u\sbar + \u\d_0$, $1/6$ into $\u\sbar + \u\d_1$ and $1/3$ into $\d\sbar + \u\u_1$. Once the flavours of the meson are determined, the choice of meson multiplet is performed as in the standard fragmentation description. \item Finally, an antiquark initiator, such as an $\sbar$, leaves behind a $\u\u\d\s$ four-quark state, which is subdivided into a baryon plus a quark. Since, to first approximation, the $\s\sbar$ pair comes from the branching $\g \to \s\sbar$ of a colour octet gluon, the subdivision $\u\u\d + \s$ is not allowed, since it would correspond to a colour-singlet $\s\sbar$. Therefore the subdivision is $1/2$ into $\u\d_0\s + \u$, $1/6$ into $\u\d_1\s + \u$ and $1/3$ into $\u\u_1\s + \d$. A baryon is formed among the ones possible for the given flavour content and diquark spin, according to the relative probabilities used in the fragmentation. One could argue for an additional weighting to count the number of baryon states available for a given diquark plus quark combination, but this has not been included. \end{Itemize} One may note that any $\u$ or $\d$ quark taken out of the proton is automatically assumed to be a valence quark. Clearly this is unrealistic, but not quite as bad as it might seem. In particular, one should remember that the beam remnant scenario is applied to the initial-state shower initiators at a scale of $Q_0 \approx 1$ GeV and at an $x$ value usually much larger than the $x$ at the hard scattering. The sea quark contribution therefore normally is negligible. For a meson beam remnant, the rules are in the same spirit, but somewhat easier, since no diquark or baryons need be taken into account. Thus a valence quark (antiquark) initiator leaves behind a valence antiquark (quark), a gluon initiator leaves behind a valence quark plus a valence antiquark, and a sea quark (antiquark) leaves behind a meson (which contains the partner to the sea parton) plus a valence antiquark (quark). A resolved photon is even simpler than a meson, since one does not have to make the distinction between valence and sea flavour. Thus any quark (antiquark) initiator leaves behind the matching antiquark (quark), and a gluon leaves behind a quark + antiquark pair. The relative quark flavour composition in the latter case is assumed proportional to $e_{\q}^2$ among light flavours, i.e. $2/3$ into $\u + \ubar$, $1/6$ into $\d + \dbar$, and $1/6$ into $\s + \sbar$. If one wanted to, one could also have chosen to represent the remnant by a single gluon. If no initial-state radiation is assumed, an electron (or, in general, a lepton or a neutrino) leaves behind no beam remnant. Also when radiation is included, one would expect to recover a single electron with the full beam energy when the shower initiator is reconstructed. This does not have to happen, e.g. if the initial-state shower is cut off at a non-vanishing scale, such that some of the emission at low $Q^2$ values is not simulated. Further, for purely technical reasons, the distribution of an electron inside an electron, $f_{\e}^{\e}(x,Q^2)$, is cut off at $x = 1 - 10^{-6}$. This means that always, when initial-state radiation is included, a fraction of at least $10^{-6}$ of the beam energy has to be put into one single photon along the beam direction, to represent this not simulated radiation. The physics is here slightly different from the standard beam remnant concept, but it is handled with the same machinery. Beam remnants can also apper when the electron is resolved with the use of parton distributions, but initial-state radiation is switched off. Conceptually, this is a contradiction, since it is the initial-state radiation that builds up the parton distributions, but sometimes the combination is still useful. Finally, since QED radiation has not yet been included in events with resolved photons inside electrons, also in this case effective beam remnants have to be assigned by the program. The beam remnant assignments inside an electron, in either of the cases above, is as follows. \begin{Itemize} \item An $\e^-$ initiator leaves behind a $\gamma$ remnant. \item A $\gamma$ initiator leaves behind an $\e^-$ remnant. \item An $\e^+$ initiator leaves behind an $\e^- + \e^-$ remnant. \item A $\q$ ($\qbar$) initiator leaves behind a $\qbar + \e^-$ ($\q + \e^-$) remnant. \item A $\g$ initiator leaves behind a $\g + \e^-$ remnant. One could argue that, in agreement with the treatment of photon beams above, the remnant should be $\q + \qbar + \e^-$. The program currently does not allow for three beam remnant objects, however. \end{Itemize} By the hard scattering and initial-state radiation machinery, the shower initiator has been assigned some fraction $x$ of the four-momentum of the beam particle, leaving behind $1-x$ to the remnant. If the remnant consists of two objects, this energy and momentum has to be shared, somehow. For an electron, the sharing is given from first principles: if, e.g., the initiator is a $\q$, then that $\q$ was produced in the sequence of branchings $\e \to \gamma \to \q$, where $x_{\gamma}$ is distributed according to the convolution in eq. (\ref{pg:foldqgine}). Therefore the $\qbar$ remnant takes a fraction $\chi = (x_{\gamma} -x)/(1-x)$ of the total remnant energy, and the $\e$ takes $1 - \chi$. For the other beam remnants, the relative energy-sharing variable $\chi$ is not known from first principles, but picked according to some suitable parametrization. Normally several different options are available, that can be set separately for baryon and meson beams, and for hadron + quark and quark + diquark (or antiquark) remnants. In one extreme are shapes in agreement with na\"{\i}ve counting rules, i.e. where energy is shared evenly between `valence' partons. For instance, ${\cal P}(\chi) = 2 \, (1-\chi)$ for the energy fraction taken by the $\q$ in a $\q + \q\q$ remnant. In the other extreme, an uneven distribution could be used, like in parton distributions, where the quark only takes a small fraction and most is retained by the diquark. The default for a $\q + \q\q$ remnant is of this type, \begin{equation} {\cal P}(\chi) \propto \frac{(1 - \chi)^3} {\sqrt{\chi^2 + c_{\mmin}^2}} ~, \end{equation} with $c_{\mmin} = 2 \langle m_{\q} \rangle / E_{\mrm{cm}} = (0.6$ GeV)$/ E_{\mrm{cm}}$ providing a lower cut-off. In general, the more uneven the sharing of the energy, the less the total multiplicity in the beam remnant fragmentation. If no multiple interactions are allowed, a rather even sharing is needed to come close to the experimental multiplicity (and yet one does not quite make it). With an uneven sharing there is room to generate more of the total multiplicity by multiple interactions \cite{Sjo87a}. In a photon beam, with a remnant $\q + \qbar$, the $\chi$ variable is chosen the same way it would have been in a corresponding meson remnant. Before the $\chi$ variable is used to assign remnant momenta, it is also necessary to consider the issue of primordial $k_{\perp}$. The initiator partons are thus assigned each a $k_{\perp}$ value, vanishing for an electron or photon inside an electron, distributed either according to a Gaussian or an exponential shape for a hadron, and according to either of these shapes or a power-like shape for a quark or gluon inside a photon (which may in its turn be inside an electron). The interaction subsystem is boosted and rotated to bring it from the frame assumed so far, with each initiator along the $\pm z$ axis, to one where the initiators have the required primordial $k_{\perp}$ values. The $\pT$ recoil is taken by the remnant. If the remnant is composite, the recoil is all taken by one of the two, namely the one that, in some imagined perturbative splitting language, is the sister of the initiator parton. For instance, when a gluon is taken out of a proton, the recoil is all taken by the lone quark (i.e. nothing by the diquark), since one could have imagined an earlier branching $\q_0 \to \q \g$, below the shower cut-off scale $Q_0$, with $p_{\perp q_0} = 0$. In addition, however, two beam remnants may be given a relative $\pT$, which is then always chosen as for $\q_i \qbar_i$ pairs in the fragmentation description. The $\chi$ variable is interpreted as a sharing of light-cone energy and momentum, i.e. $E + p_z$ for the beam moving in the $+z$ direction and $E - p_z$ for the other one. When the two transverse masses $m_{\perp 1}$ and $m_{\perp 2}$ of a composite remnant have been constructed, the total transverse mass can therefore be found as \begin{equation} m_{\perp}^2 = \frac{m_{\perp 1}^2}{\chi} + \frac{m_{\perp 2}^2}{1 - \chi} ~, \end{equation} if remnant 1 is the one that takes the fraction $\chi$. The choice of a light-cone interpretation to $\chi$ means the definition is invariant under longitudinal boosts, and therefore does not depend on the beam energy itself. A $\chi$ value close to the na\"{\i}ve borders 0 or 1 can lead to an unreasonably large remnant $m_{\perp}$. Therefore an additional check is introduced, that the remnant $m_{\perp}$ be smaller than the na\"{\i}ve c.m. frame remnant energy, $(1-x) E_{\mrm{cm}}/2$. If this is not the case, a new $\chi$ and a new relative transverse momentum is selected. Whether there is one remnant parton or two, the transverse mass of the remnant is not likely to agree with $1-x$ times the mass of the beam particle, i.e. it is not going to be possible to preserve the energy and momentum in each remnant separately. One therefore allows a shuffling of energy and momentum between the beam remnants from each of the two incoming beams. This may be achieved by performing a (small) longitudinal boost of each remnant system. Since there are two boost degrees of freedom, one for each remnant, and two constraints, one for energy and one for longitudinal momentum, a solution may be found. Under some circumstances, one beam remnant may be absent or of very low energy, while the other one is more complicated. One example is deep inelastic scattering in $\ep$ collisions, where the electron leaves no remnant, or maybe only a low-energy photon. It is clearly then not possible to balance the two beam remnants against each other. Therefore, if one beam remnant has an energy below 0.2 of the beam energy, i.e. if the initiator parton has $x > 0.8$, then the two boosts needed to ensure energy and momentum conservation are instead performed on the other remnant and on the interaction subsystem. If there is a low-energy remnant at all then, before that, energy and momentum are assigned to the remnant constituent(s) so that the appropriate light-cone combination $E \pm p_z$ is conserved, but not energy or momentum separately. If both beam remnants have low energy, but both still exist, then the one with lower $m_{\perp} / E$ is the one that will not be boosted. \subsection{Multiple Interactions} \label{ss:multint} In this section we present the model used in {\Py} to describe the possibility that several parton pairs undergo hard interactions in a hadron--hadron collision, and thereby contribute to the overall event activity, in particular at low $\pT$. The same model is also used to describe the VMD $\gamma \p$ events, where the photon interacts like a hadron. It should from the onset be made clear that this is not an easy topic. In fact, in the full event generation process, probably no other area is as poorly understood as this one. The whole concept of multiple interactions is very controversial, with contraditory experimental conclusions \cite{AFS87}. The multiple interactions scenario presented here \cite{Sjo87a} was the first detailed model for this kind of physics, and is still one of the very few available. We will present two related but separate scenarios, one `simple' model and one somewhat more sophisticated. In fact, neither of them are all that simple, which may make the models look unattractive. However, the world of hadron physics {\it is} complicated, and if we err, it is most likely in being too unsophisticated. The experience gained with the model(s), in failures as well as successes, could be used as a guideline in the evolution of yet more detailed models. Our basic philosophy will be as follows. The total rate of parton--parton interactions, as a function of the transverse momentum scale $\pT$, is assumed to be given by perturbative QCD. This is certainly true for reasonably large $\pT$ values, but here we shall also extend the perturbative parton--parton scattering framework into the low-$\pT$ region. A regularization of the divergence in the cross section for $\pT \to 0$ has to be introduced, however, which will provide us with the main free parameter of the model. Since each incoming hadron is a composite object, consisting of many partons, there should exist the possibility of several parton pairs interacting when two hadrons collide. It is not unreasonable to assume that the different pairwise interactions take place essentially independently of each other, and that therefore the number of interactions in an event is given by a Poissonian distribution. This is the strategy of the `simple' scenario. Furthermore, hadrons are not only composite but also extended objects, meaning that collisions range from very central to rather peripheral ones. Reasonably, the average number of interactions should be larger in the former than in the latter case. Whereas the assumption of a Poissonian distribution should hold for each impact parameter separately, the distribution in number of interactions should be widened by the spread of impact parameters. The amount of widening depends on the assumed matter distribution inside the colliding hadrons. In the `complex' scenario, different matter distributions are therefore introduced. \subsubsection{The basic cross sections} The QCD cross section for hard $2 \to 2$ processes, as a function of the $\pT^2$ scale, is given by \begin{equation} \frac{\d \sigma}{\d \pT^2} = \sum_{i,j,k} \int \d x_1 \int \d x_2 \int \d \hat{t} \, f_i(x_1,Q^2) \, f_j(x_2,Q^2) \, \frac{\d \hat{\sigma}_{ij}^k}{\d \hat{t}} \, \delta \! \left( \pT^2 - \frac{\hat{t}\hat{u}}{\hat{s}} \right) ~, \label{mi:sigmapt} \end{equation} cf. section \ref{ss:kinemtwo}. Implicitly, from now on we are assuming that the `hardness' of processes is given by the $\pT$ scale of the scattering. For an application of the formula above to small $\pT$ values, a number of caveats could be made. At low $\pT$, the integrals receive major contributions from the small-$x$ region, where parton distributions are poorly understood theoretically (Regge limit behaviour, dense packing problems etc. \cite{Lev90}) and not yet measured. Different sets of parton distributions can therefore give numerically rather different results for the phenomenology of interest. One may also worry about higher-order corrections to the jet rates ($K$ factors), beyond what is given by parton-shower corrections --- one simple option we allow here is to evaluate $\alphas$ of the hard scattering process at an optimized scale, $\alphas(0.075\pT^2)$ \cite{Ell86}. The hard scattering cross section above some given $\pTmin$ is given by \begin{equation} \sigma_{\mrm{hard}} (\pTmin) = \int_{\pTmin^2}^{s/4} \frac{\d \sigma}{\d \pT^2} \, \d \pT^2 ~. \label{mi:sigmahard} \end{equation} Since the differential cross section diverges roughly like $\d \pT^2 / \pT^4$, $\sigma_{\mrm{hard}}$ is also divergent for $\pTmin \to 0$. We may compare this with the total inelastic, non-diffractive cross section $\sigma_{\mrm{nd}}(s)$ --- elastic and diffractive events are not the topic of this section. At current collider energies $\sigma_{\mrm{hard}} (\pTmin)$ becomes comparable with $\sigma_{\mrm{nd}}$ for $\pTmin \approx$1.5--2 GeV. This need not lead to contradictions: $\sigma_{\mrm{hard}}$ does not give the hadron--hadron cross section but the parton--parton one. Each of the incoming hadrons may be viewed as a beam of partons, with the possibility of having several parton--parton interactions when the hadrons pass through each other. In this language, $\sigma_{\mrm{hard}} (\pTmin) / \sigma_{\mrm{nd}}(s)$ is simply the average number of parton--parton scatterings above $\pTmin$ in an event, and this number may well be larger than unity. While the introduction of several interactions per event is the natural consequence of allowing small $\pTmin$ values and hence large $\sigma_{\mrm{hard}}$ ones, it is not the solution of $\sigma_{\mrm{hard}} (\pTmin)$ being divergent for $\pTmin \to 0$: the average $\hat{s}$ of a scattering decreases slower with $\pTmin$ than the number of interactions increases, so na\"{\i}vely the total amount of scattered partonic energy becomes infinite. One cut-off is therefore obtained via the need to introduce proper multi-parton correlated parton distributions inside a hadron. This is not a part of the standard perturbative QCD formalism and is therefore not built into eq. (\ref{mi:sigmahard}). In practice, even correlated parton-distribution functions seems to provide too weak a cut, i.e. one is lead to a picture with too little of the incoming energy remaining in the small-angle beam jet region. A more credible reason for an effective cut-off is that the incoming hadrons are colour neutral objects. Therefore, when the $\pT$ of an exchanged gluon is made small and the transverse wavelength correspondingly large, the gluon can no longer resolve the individual colour charges, and the effective coupling is decreased. This mechanism is not in contradiction to perturbative QCD calculations, which are always performed assuming scattering of free partons (rather than partons inside hadrons), but neither does present knowledge of QCD provide an understanding of how such a decoupling mechanism would work in detail. In the simple model one makes use of a sharp cut-off at some scale $\pTmin$, while a more smooth dampening is assumed for the complex scenario. \subsubsection{The simple model} In an event with several interactions, it is convenient to impose an ordering. The logical choice is to arrange the scatterings in falling sequence of $x_{\perp} = 2 \pT / E_{\mrm{cm}}$. The `first' scattering is thus the hardest one, with the `subsequent' (`second', `third', etc.) successively softer. It is important to remember that this terminology is in no way related to any picture in physical time; we do not know anything about the latter. In principle, all the scatterings that occur in an event must be correlated somehow, na\"{\i}vely by momentum and flavour conservation for the partons from each incoming hadron, less na\"{\i}vely by various quantum mechanical effects. When averaging over all configurations of soft partons, however, one should effectively obtain the standard QCD phenomenology for a hard scattering, e.g. in terms of parton distributions. Correlation effects, known or estimated, can be introduced in the choice of subsequent scatterings, given that the `preceding' (harder) ones are already known. With a total cross section of hard interactions $\sigma_{\mrm{hard}} (\pTmin)$ to be distributed among $\sigma_{\mrm{nd}}(s)$ (non-diffractive, inelastic) events, the average number of interactions per event is just the ratio $\br{n} = \sigma_{\mrm{hard}} (\pTmin) / \sigma_{\mrm{nd}}(s)$. As a starting point we will assume that all hadron collisions are equivalent (no impact parameter dependence), and that the different parton--parton interactions take place completely independently of each other. The number of scatterings per event is then distributed according to a Poissonian with mean $\br{n}$. A fit to collider multiplicity data gives $\pTmin \approx 1.6$ GeV, which corresponds to $\br{n} \approx 1$. For Monte Carlo generation of these interactions it is useful to define \begin{equation} f(x_{\perp}) = \frac{1}{\sigma_{\mrm{nd}}(s)} \, \frac{\d \sigma}{\d x_{\perp}} ~, \end{equation} with $\d \sigma / \d x_{\perp}$ obtained by analogy with eq. (\ref{mi:sigmapt}). Then $f(x_{\perp})$ is simply the probability to have a parton--parton interaction at $x_{\perp}$, given that the two hadrons undergo a non-diffractive, inelastic collision. The probability that the hardest interaction, i.e. the one with highest $x_{\perp}$, is at $x_{\perp 1}$, is now given by \begin{equation} f(x_{\perp 1}) \exp \left\{ - \int_{x_{\perp 1}}^1 f(x'_{\perp}) \, \d x'_{\perp} \right\} ~, \label{mi:hardest} \end{equation} i.e. the na\"{\i}ve probability to have a scattering at $x_{\perp 1}$ multiplied by the probability that there was no scattering with $x_{\perp}$ larger than $x_{\perp 1}$. This is the familiar exponential dampening in radioactive decays, encountered e.g. in parton showers in section \ref{ss:sudakov}. Using the same technique as in the proof of the veto algorithm, section \ref{ss:vetoalg}, the probability to have an $i$:th scattering at an $x_{\perp i} < x_{\perp i-1} < \cdots < x_{\perp 1} < 1$ is found to be \begin{equation} f(x_{\perp i}) \, \frac{1}{(i-1)!} \left( \int_{x_{\perp i}}^1 f(x'_{\perp}) \, \d x'_{\perp} \right)^{i-1} \exp \left\{ - \int_{x_{\perp i}}^1 f(x'_{\perp}) \, \d x'_{\perp} \right\} ~. \end{equation} The total probability to have a scattering at a given $x_{\perp}$, irrespectively of it being the first, the second or whatever, obviously adds up to give back $f(x_{\perp})$. The multiple interaction formalism thus retains the correct perturbative QCD expression for the scattering probability at any given $x_{\perp}$. With the help of the integral \begin{equation} F(x_{\perp}) = \int_{x_{\perp}}^1 f(x'_{\perp}) \, \d x'_{\perp} = \frac{1}{\sigma_{\mrm{nd}}(s)} \, \int_{s x_{\perp}^2/4}^{s/4} \frac{\d \sigma}{\d \pT^2} \, \d \pT^2 \end{equation} (where we assume $F(x_{\perp}) \to \infty$ for $x_{\perp} \to 0$) and its inverse $F^{-1}$, the iterative procedure to generate a chain of scatterings $1 > x_{\perp 1} > x_{\perp 2} > \cdots > x_{\perp i}$ is given by \begin{equation} x_{\perp i} = F^{-1}(F(x_{\perp i-1}) - \ln R_i) ~. \label{mi:iter} \end{equation} Here the $R_i$ are random numbers evenly distributed between 0 and 1. The iterative chain is started with a fictitious $x_{\perp 0} = 1 $ and is terminated when $x_{\perp i}$ is smaller than $x_{\perp \mmin} = 2 \pTmin / E_{\mrm{cm}}$. Since $F$ and $F^{-1}$ are not known analytically, the standard veto algorithm is used to generate a much denser set of $x_{\perp}$ values, whereof only some are retained in the end. In addition to the $\pT^2$ of an interaction, it is also necessary to generate the other flavour and kinematics variables according to the relevant matrix elements. Whereas the ordinary parton distributions should be used for the hardest scattering, in order to reproduce standard QCD phenomenology, the parton distributions to be used for subsequent scatterings must depend on all preceding $x$ values and flavours chosen. We do not know enough about the hadron wave function to write down such joint probability distributions. To take into account the energy `already' used in harder scatterings, a conservative approach is to evaluate the parton distributions, not at $x_i$ for the $i$:th scattered parton from hadron, but at the rescaled value \begin{equation} x'_i = \frac{x_i}{\sum_{j=1}^{i-1} x_j} ~. \end{equation} This is our standard procedure in the program; we have tried a few alternatives without finding any significantly different behaviour in the final physics. In a fraction $\exp(-F(x_{\perp \mmin}))$ of the events studied, there will be no hard scattering above $x_{\perp \mmin}$ when the iterative procedure in eq. (\ref{mi:iter}) is applied. It is therefore also necessary to have a model for what happens in events with no (semi)hard interactions. The simplest possible way to produce an event is to have an exchange of a very soft gluon between the two colliding hadrons. Without (initially) affecting the momentum distribution of partons, the `hadrons' become colour octet objects rather than colour singlet ones. If only valence quarks are considered, the colour octet state of a baryon can be decomposed into a colour triplet quark and an antitriplet diquark. In a baryon-baryon collision, one would then obtain a two-string picture, with each string stretched from the quark of one baryon to the diquark of the other. A baryon-antibaryon collision would give one string between a quark and an antiquark and another one between a diquark and an antidiquark. In a hard interaction, the number of possible string drawings are many more, and the overall situation can become quite complex when several hard scatterings are present in an event. Specifically, the string drawing now depends on the relative colour arrangement, in each hadron individually, of the partons that are about to scatter. This is a subject about which nothing is known. To make matters worse, the standard string fragmentation description would have to be extended, to handle events where two or more valence quarks have been kicked out of an incoming hadron by separate interactions. In particular, the position of the baryon number would be unclear. We therefore here assume that, following the hardest interaction, all subsequent interactions belong to one of three classes. \begin{Itemize} \item Scatterings of the $\g\g \to \g\g $ type, with the two gluons in a colour-singlet state, such that a double string is stretched directly between the two outgoing gluons, decoupled from the rest of the system. \item Scatterings $\g\g \to \g\g$, but colour correlations assumed to be such that each of the gluons is connected to one of the strings `already' present. Among the different possibilities of connecting the colours of the gluons, the one which minimizes the total increase in string length is chosen. This is in contrast to the previous alternative, which roughly corresponds to a maximization of the extra string length. \item Scatterings $\g\g \to \q\qbar$, with the final pair again in a colour-singlet state, such that a single string is stretched between the outgoing $\q$ and $\qbar$. \end{Itemize} By default, the three possibilities are assumed equally probable. Note that the total jet rate is maintained at its nominal value, i.e. scatterings such as $\q\g \to \q\g$ are included in the cross section, but are replaced by a mixture of $\g\g$ and $\q\qbar$ events for string drawing issues. Only the hardest interaction is guaranteed to give strings coupled to the beam remnants. One should not take this approach to colour flow too seriously --- clearly it is a simplification --- but the overall picture does not tend to be very dependent on the particular choice you make. Since a $\g\g \to \g\g$ or $\q\qbar$ scattering need not remain that if initial- and final-state showers were to be included, radiation is only included for the hardest interaction. In practice, there is no problem: except for the hardest interaction, which can be hard because of experimental trigger conditions, it is unlikely for a parton scattering to be so hard that radiation plays a significant r\^ole. In events with multiple interactions, the beam remnant treatment is slightly modified. First the hard scattering is generated, with its associated initial- and final-state radiation, and next any additional multiple interactions. Only thereafter are beam remnants attached to the initator partons of the hardest scattering, using the same machinery as before, except that the energy and momentum already taken away from the beam remnants also include that of the subsequent interactions. \subsubsection{A model with varying impact parameters} Up to this point, it has been assumed that the initial state is the same for all hadron collisions, whereas in fact each collision also is characterized by a varying impact parameter $b$. Within the classical framework of this paper, $b$ is to be thought of as a distance of closest approach, not as the Fourier transform of the momentum transfer. A small $b$ value corresponds to a large overlap between the two colliding hadrons, and hence an enhanced probability for multiple interactions. A large $b$, on the other hand, corresponds to a grazing collision, with a large probability that no parton--parton interactions at all take place. In order to quantify the concept of hadronic matter overlap, one may assume a spherically symmetric distribution of matter inside the hadron, $\rho(\mbf{x}) \, \d^3 x = \rho(r) \, \d^3 x$. For simplicity, the same spatial distribution is taken to apply for all parton species and momenta. Several different matter distributions have been tried, and are available. We will here concentrate on the most extreme one, a double Gaussian \begin{equation} \rho(r) \propto \frac{1 - \beta}{a_1^3} \, \exp \left\{ - \frac{r^2}{a_1^2} \right\} + \frac{\beta}{a_2^3} \, \exp \left\{ - \frac{r^2}{a_2^2} \right\} ~. \label{mi:doubleGauss} \end{equation} This corresponds to a distribution with a small core region, of radius $a_2$ and containing a fraction $\beta$ of the total hadronic matter, embedded in a larger hadron of radius $a_1$. While it is mathematically convenient to have the origin of the two Gaussians coinciding, the physics could well correspond to having three disjoint core regions, reflecting the presence of three valence quarks, together carrying the fraction $\beta$ of the proton momentum. One could alternatively imagine a hard hadronic core surrounded by a pion cloud. Such details would affect e.g. the predictions for the $t$ distribution in elastic scattering, but are not of any consequence for the current topics. To be specific, the values $\beta = 0.5$ and $a_2/a_1 = 0.2$ have been picked as default values. It should be noted that the overall distance scale $a_1$ never enters in the subsequent calculations, since the inelastic, non-diffractive cross section $\sigma_{\mrm{nd}}(s)$ is taken from literature rather than calculated from the $\rho(r)$. Compared to other shapes, like a simple Gaussian, the double Gaussian tends to give larger fluctuations, e.g. in the multiplicity distribution of minimum bias events: a collision in which the two cores overlap tends to have a strongly increased activity, while ones where they do not are rather less active. One also has a biasing effect: hard processes are more likely when the cores overlap, thus hard scatterings are associated with an enhanced multiple interaction rate. This provides one possible explanation for the experimental `pedestal effect'. For a collision with impact parameter $b$, the time-integrated overlap ${\cal O}(b)$ between the matter distributions of the colliding hadrons is given by \begin{eqnarray} & & {\cal O}(b) \propto \int \d t \int \d^3 x \, \rho(x,y,z) \, \rho(x+b,y,z+t) \nonumber \\ & & \propto \frac{(1 - \beta)^2}{2a_1^2} \exp \left\{ - \frac{b^2}{2a_1^2} \right\} + \frac{2 \beta (1-\beta)}{a_1^2+a_2^2} \exp \left\{ - \frac{b^2}{a_1^2+ a_2^2} \right\} + \frac{\beta^2}{2a_2^2} \exp \left\{ - \frac{b^2}{2a_2^2} \right\} ~. \end{eqnarray} The necessity to use boosted $\rho(\mbf{x})$ distributions has been circumvented by a suitable scale transformation of the $z$ and $t$ coordinates. The overlap ${\cal O}(b)$ is obviously strongly related to the eikonal $\Omega(b)$ of optical models. We have kept a separate notation, since the physics context of the two is slightly different: $\Omega(b)$ is based on the quantum mechanical scattering of waves in a potential, and is normally used to describe the elastic scattering of a hadron-as-a-whole, while ${\cal O}(b)$ comes from a purely classical picture of point-like partons distributed inside the two colliding hadrons. Furthermore, the normalization and energy dependence is differently realized in the two formalisms. The larger the overlap ${\cal O}(b)$ is, the more likely it is to have interactions between partons in the two colliding hadrons. In fact, there should be a linear relationship \begin{equation} \langle \tilde{n}(b) \rangle = k {\cal O}(b) ~, \end{equation} where $\tilde{n} = 0, 1, 2, \ldots$ counts the number of interactions when two hadrons pass each other with an impact parameter $b$. The constant of proportionality, $k$, is related to the parton--parton cross section and hence increases with c.m. energy. For each given impact parameter, the number of interactions is assumed to be distributed according to a Poissonian. If the matter distribution has a tail to infinity (as the double Gaussian does), events may be obtained with arbitrarily large $b$ values. In order to obtain finite total cross sections, it is necessary to assume that each event contains at least one semi-hard interaction. The probability that two hadrons, passing each other with an impact parameter $b$, will actually undergo a collision is then given by \begin{equation} {\cal P}_{\mrm{int}}(b) = 1 - \exp ( - \langle \tilde{n}(b) \rangle ) = 1 - \exp (- k {\cal O}(b) ) ~, \end{equation} according to Poissonian statistics. The average number of interactions per event at impact parameter $b$ is now \begin{equation} \langle n(b) \rangle = \frac{ \langle \tilde{n}(b) \rangle }{{\cal P}_{\mrm{int}}(b)} = \frac{ k {\cal O}(b) }{ 1 - \exp (- k {\cal O}(b) ) } ~, \label{mi:nofb} \end{equation} where the denominator comes from the removal of hadron pairs which pass without colliding, i.e. with $\tilde{n} = 0$. The relationship $\langle n \rangle = \sigma_{\mrm{hard}}/\sigma_{\mrm{nd}}$ was earlier introduced for the average number of interactions per non-diffractive, inelastic event. When averaged over all impact parameters, this relation must still hold true: the introduction of variable impact parameters may give more interactions in some events and less in others, but it does not affect either $\sigma_{\mrm{hard}}$ or $\sigma_{\mrm{nd}}$. For the former this is because the perturbative QCD calculations only depend on the total parton flux, for the latter by construction. Integrating eq. (\ref{mi:nofb}) over $b$, one then obtains \begin{equation} \langle n \rangle = \frac{ \int \langle n(b) \rangle \, {\cal P}_{\mrm{int}}(b) \, \d^2 b} { \int {\cal P}_{\mrm{int}}(b) \, \d^2 b} = \frac{ \int k {\cal O}(b) \, \d^2 b} { \int \left( 1 - \exp (- k {\cal O}(b) ) \right) \, \d^2 b} = \frac{\sigma_{\mrm{hard}}}{\sigma_{\mrm{nd}}} ~. \label{mi:kfromsigma} \end{equation} For ${\cal O}(b)$, $\sigma_{\mrm{hard}}$ and $\sigma_{\mrm{nd}}$ given, with $\sigma_{\mrm{hard}} / \sigma_{\mrm{nd}} > 1$, $k$ can thus always be found (numerically) by solving the last equality. The absolute normalization of ${\cal O}(b)$ is not interesting in itself, but only the relative variation with impact parameter. It is therefore useful to introduce an `enhancement factor' $e(b)$, which gauges how the interaction probability for a passage with impact parameter $b$ compares with the average, i.e. \begin{equation} \langle \tilde{n}(b) \rangle = k{\cal O}(b) = e(b) \, \langle k{\cal O}(b) \rangle ~. \label{mi:ebenh} \end{equation} The definition of the average $\langle k{\cal O}(b) \rangle$ is a bit delicate, since the average number of interactions per event is pushed up by the requirement that each event contain at least one interaction. However, an exact meaning can be given \cite{Sjo87a}. With the knowledge of $e(b)$, the $f(x_{\perp})$ function of the simple model generalizes to \begin{equation} f(x_{\perp},b) = e(b) \, f(x_{\perp}) ~. \end{equation} The na\"{\i}ve generation procedure is thus to pick a $b$ according to the phase space $\d^2 b$, find the relevant $e(b)$ and plug in the resulting $f(x_{\perp},b)$ in the formalism of the simple model. If at least one hard interaction is generated, the event is retained, else a new $b$ is to be found. This algorithm would work fine for hadronic matter distributions which vanish outside some radius, so that the $\d^2 b$ phase space which needs to be probed is finite. Since this is not true for the distributions under study, it is necessary to do better. By analogy with eq. (\ref{mi:hardest}), it is possible to ask what the probability is to find the hardest scattering of an event at $x_{\perp 1}$. For each impact parameter separately, the probability to have an interaction at $x_{\perp 1}$ is given by $f(x_{\perp},b)$, and this should be multiplied by the probability that the event contains no interactions at a scale $x'_{\perp} > x_{\perp 1}$, to yield the total probability distribution \begin{eqnarray} \frac{\d {\cal P}_{\mrm{hardest}}}{\d^2 b \, \d x_{\perp 1}} & = & f(x_{\perp 1},b) \, \exp \left\{ - \int_{x_{\perp 1}}^1 f(x'_{\perp},b) \, \d x'_{\perp} \right\} \nonumber \\ & = & e(b) \, f(x_{\perp 1}) \, \exp \left\{ - e(b) \int_{x_{\perp 1}}^1 f(x'_{\perp}) \, \d x'_{\perp} \right\} ~. \end{eqnarray} If the treatment of the exponential is deferred for a moment, the distribution in $b$ and $x_{\perp 1}$ appears in factorized form, so that the two can be chosen independently of each other. In particular, a high-$\pT$ QCD scattering or any other hard scattering can be selected with whatever kinematics desired for that process, and thereafter assigned some suitable `hardness' $x_{\perp 1}$. With the $b$ chosen according to $e(b) \, \d^2 b$, the neglected exponential can now be evaluated, and the event retained with a probability proportional to it. From the $x_{\perp 1}$ scale of the selected interaction, a sequence of softer $x_{\perp i}$ values may again be generated as in the simple model, using the known $f(x_{\perp},b)$. This sequence may be empty, i.e. the event need not contain any further interactions. It is interesting to understand how the algorithm above works. By selecting $b$ according to $e(b) \, \d^2 b$, i.e. ${\cal O}(b) \, \d^2 b$, the primary $b$ distribution is maximally biased towards small impact parameters. If the first interaction is hard, by choice or by chance, the integral of the cross section above $x_{\perp 1}$ is small, and the exponential close to unity. The rejection procedure is therefore very efficient for all standard hard processes in the program --- one may even safely drop the weighting with the exponential completely. The large $e(b)$ value is also likely to lead to the generation of many further, softer interactions. If, on the other hand, the first interaction is not hard, the exponential is no longer close to unity, and many events are rejected. This pulls down the efficiency for `minimum bias' event generation. Since the exponent is proportional to $e(b)$, a large $e(b)$ leads to an enhanced probability for rejection, whereas the chance of acceptance is larger with a small $e(b)$. Among events where the hardest interaction is soft, the $b$ distribution is therefore biased towards larger values (smaller $e(b)$), and there is a small probability for yet softer interactions. To evaluate the exponential factor, the program pretabulates the integral of $f(x_{\perp})$ at the initialization stage, and further increases the Monte Carlo statistics of this tabulation as the run proceeds. The $x_{\perp}$ grid is concentrated towards small $x_{\perp}$, where the integral is large. For a selected $x_{\perp 1}$ value, the $f(x_{\perp})$ integral is obtained by interpolation. After multiplication by the known $e(b)$ factor, the exponential factor may be found. In this section, nothing has yet been assumed about the form of the $\d \sigma / \d \pT$ spectrum. Like in the impact parameter independent case, it is possible to use a sharp cut-off at some given $\pTmin$ value. However, now each event is required to have at least one interaction, whereas before events without interactions were retained and put at $\pT = 0$. It is therefore aesthetically more appealing to assume a gradual turn-off, so that a (semi)hard interaction can be rather soft part of the time. The matrix elements roughly diverge like $\alphas(\pT^2) \, \d \pT^2 / \pT^4$ for $\pT \to 0$. They could therefore be regularized as follows. Firstly, to remove the $1/\pT^4$ behaviour, multiply by a factor $\pT^4 / (\pT^2 + p_{\perp 0}^2)^2$. Secondly, replace the $\pT^2$ argument in $\alphas$ by $\pT^2 + p_{\perp 0}^2$ or, with the inclusion of the $K$ factor procedure introduced earlier, replace $0.075 \, \pT^2$ by $0.075 \, (\pT^2 + p_{\perp 0}^2)$. With these substitutions, a continuous $\pT$ spectrum is obtained, stretching from $\pT = 0$ to $E_{\mrm{cm}}/2$. For $\pT \gg p_{\perp 0}$ the standard perturbative QCD cross section is recovered, while values $\pT \ll p_{\perp 0}$ are strongly damped. The $p_{\perp 0}$ scale, which now is the main free parameter of the model, in practice comes out to be of the same order of magnitude as the sharp cut-off $\pTmin$ did, i.e. 1.5--2 GeV. If gluons with large transverse wavelength decouple because of the colour-singlet nature of hadrons, and if the transverse structure of hadrons is assumed to be energy-independent, it is natural to assume that also $\pTmin$ and $p_{\perp 0}$ are independent of the c.m. energy of the hadron collision. For the impact parameter independent picture this works out fine, with all events being reduced to low-$\pT$ two-string ones when the c.m. energy is reduced. In the variable impact parameter picture, the whole formalism only makes sense if $\sigma_{\mrm{hard}} > \sigma_{\mrm{nd}}$, see e.g. eq. (\ref{mi:kfromsigma}). Since $\sigma_{\mrm{nd}}$ does not vanish with decreasing energy, but $\sigma_{\mrm{hard}}$ would do that for a fixed $p_{\perp 0}$, this means that $p_{\perp 0}$ has to be reduced when the energy is decreased below some given threshold. The more `sophisticated' model of this section therefore makes sense at collider energies, whereas it is not well suited for applications at lower energies. \subsection{Pile-up Events} \label{ss:pileup} In high luminosity colliders, there is a non-negligible probability that one single bunch crossing may produce several separate events, so-called pile-up events. This in particular applies to future $\p\p$ colliders like LHC, but one could also consider e.g. $\ee$ colliders with high rates of $\gamma\gamma$ collisions. The program therefore contains an option, currently only applicable to hadron--hadron collisions, wherein several events may be generated and put one after the other in the event record, to simulate the full amount of particle production a detector might be facing. The program needs to know the assumed luminosity per bunch--bunch crossing, expressed in mb$^{-1}$. Multiplied by the cross section for pile-up processes studied, $\sigma_{\mrm{pile}}$, this gives the average number of collisions per beam crossing, $\br{n}$. These pile-up events are taken to be of the minimum bias type, with diffractive and elastic events included or not (and a further subdivision of diffractive events into single and double). This means that $\sigma_{\mrm{pile}}$ may be either $\sigma_{\mrm{tot}}$, $\sigma_{\mrm{tot}} - \sigma_{\mrm{el}}$ or $\sigma_{\mrm{tot}} - \sigma_{\mrm{el}} - \sigma_{\mrm{diffr}}$. Which option to choose depends on the detector: most detectors would not be able to observe elastic $\p\p$ scattering, and therefore it would be superfluous to generate that kind of events. In addition, we allow for the possibility that one interaction may be of a rare kind, selected freely by the user. There is no option to generate two `rare' events in the same crossing; normally the likelihood for that kind of occurences should be small. If only minimum bias type events are generated, i.e. if only one cross section is involved in the problem, then the number of events in a crossing is distributed according to a Poissonian with the average number $\br{n}$ as calculated above. The program actually will simulate only those beam crossings where at least one event occurs, i.e. not consider the fraction $\exp(-\br{n})$ of zero-event crossings. Therefore the actually generated average number of pile-up events is $\langle n \rangle = \br{n}/(1-\exp(-\br{n}))$. Now instead consider the other extreme, where one event is supposed be rare, with a cross section $\sigma_{\mrm{rare}}$ much smaller than $\sigma_{\mrm{pile}}$, i.e. $f \equiv \sigma_{\mrm{rare}}/\sigma_{\mrm{pile}} \ll 1$. The probability that a bunch crossing will give $i$ events, whereof one of the rare kind, now is \begin{equation} {\cal P}_i = f \, i \, \exp(-\br{n}) \, \frac{\br{n}^i}{i!} = f \, \br{n} \exp(-\br{n}) \, \frac{\br{n}^{i-1}}{(i-1)!} ~. \end{equation} The na\"{\i}ve Poissonian is suppressed by a factor $f$, since one of the events is rare rather than of the normal kind, but enhanced by a factor $i$, since any one of the $i$ events may be the rare one. As the equality shows, the probability distribution is now a Poissonian in $i-1$: in a beam crossing which produces one rare event, the multiplicity of additional pile-up events is distributed according to a Poissonian with average number $\br{n}$. The total average number of events thus is $\langle n \rangle = \br{n} + 1$. Clearly, for processes with intermediate cross sections, $\br{n} \, \sigma_{\mrm{rare}}/\sigma_{\mrm{pile}} \simeq 1$, also the average number of events will be intermediate, and it is not allowed to assume only one event to be of the `rare' type. We do not consider that kind of situations. Kindly note that, in the current implementation, all events are supposed to be produced at the same vertex (the origin). To simulate the spatial extent of the colliding beams, you would have to assign interaction points yourself, and then shift each event separately by the required amount in space and time. When the pile-up option is used, one main limitation is that event records may become very large when several events are put one after the other, so that the space limit in the {\tt LUJETS} common block is reached. It is possible to expand the dimension of the common block, see {\tt MSTU(4)} and {\tt MSTU(5)}, but only up to about 20\,000 entries, which might not always be enough. For practical reasons, the program will only allow a $\br{n}$ up to 120. The multiplicity distribution is truncated above 200, or when the probability for a multiplicity has fallen below $10^{-6}$, whichever occurs sooner. Also low multiplicities with probabilities below $10^{-6}$ are truncated. \subsection{Common Block Variables} \label{ss:multintpar} Of the routines used to generate beam remnants, multiple interactions and pile-up events, none are intended to be used in standalone mode. The only way to regulate these aspects is therefore via the variables in the {\tt PYPARS} common block. \drawbox{COMMON/PYPARS/MSTP(200),PARP(200),MSTI(200),PARI(200)} \begin{entry} \itemc{Purpose:} to give access to a number of status codes and parameters which regulate the performance of {\Py}. Most parameters are described in section \ref{ss:PYswitchpar}; here only those related to beam remnants, multiple interactions and pile-up events are described. If the default values, below denoted by (D=\ldots), are not satisfactory, they must in general be changed before the {\tt PYINIT} call. Exceptions, i.e. variables which can be changed for each new event, are denoted by (C). \iteme{MSTP(81) :}\label{p:MSTP81} (D=1) master switch for multiple interactions. \begin{subentry} \iteme{= 0 :} off. \iteme{= 1 :} on. \end{subentry} \iteme{MSTP(82) :} (D=1) structure of multiple interactions. For QCD processes, used down to $\pT$ values below $\pTmin$, it also affects the choice of structure for the one hard/semi-hard interaction. \begin{subentry} \iteme{= 0 :} simple two-string model without any hard interactions. \iteme{= 1 :} multiple interactions assuming the same probability in all events, with an abrupt $\pTmin$ cut-off at {\tt PARP(81)}. \iteme{= 2 :} multiple interactions assuming the same probability in all events, with a continuous turn-off of the cross section at $p_{\perp 0} = ${\tt PARP(82)}. \iteme{= 3 :} multiple interactions assuming a varying impact parameter and a hadronic matter overlap consistent with a Gaussian matter distribution, with a continuous turn-off of the cross section at $p_{\perp 0} = ${\tt PARP(82)}. \iteme{= 4 :} multiple interactions assuming a varying impact parameter and a hadronic matter overlap consistent with a double Gaussian matter distribution given by {\tt PARP(83)} and {\tt PARP(84)}, with a continuous turn-off of the cross section at $p_{\perp 0} = ${\tt PARP(82)}. \itemc{Note 1:} For {\tt MSTP(82)}$\geq 2$ and {\tt CKIN(3)}$ > ${\tt PARP(82)}, cross sections given with {\tt PYSTAT(1)} may be somewhat too large, since (for reasons of efficiency) the probability factor that the hard interaction is indeed the hardest in the event is not included in the cross sections. It is included in the event selection, however, so the events generated are correctly distributed. For {\tt CKIN(3)} values a couple of times larger than {\tt PARP(82)} this ceases to be a problem. \itemc{Note 2:} The {\tt PARP(81)} and, in particular, {\tt PARP(82)} values are sensitive to the choice of parton distributions, $\Lambda_{\mrm{QCD}}$, etc., in the sense that a change in the latter variables leads to a net change in the multiple interaction rate, which has to be compensated by a retuning of {\tt PARP(81)} or {\tt PARP(82)} if one wants to keep the net multiple interaction structure the same. The default {\tt PARP(81)} value is consistent with the other default values give, i.e. CTEQ 2L parton distributions etc. When options {\tt MSTP(82)=} 2--4 are used, the default {\tt PARP(82)} value is to be used in conjunction with {\tt MSTP(2)=2} and {\tt MSTP(33)=3}. These switches should be set by you. \end{subentry} \iteme{MSTP(83) :} (D=100) number of Monte Carlo generated phase-space points per bin (whereof there are 20) in the initialization (in {\tt PYMULT}) of multiple interactions for {\tt MSTP(82)}$\geq 2$. \iteme{MSTP(91) :} (D=1) (C) primordial $k_{\perp}$ distribution in hadron. See {\tt MSTP(93)} for photon. \begin{subentry} \iteme{= 0 :} no primordial $k_{\perp}$. \iteme{= 1 :} Gaussian, width given in {\tt PARP(91)}, upper cut-off in {\tt PARP(93)}. \iteme{= 2 :} exponential, width given in {\tt PARP(92)}, upper cut-off in {\tt PARP(93)}. \end{subentry} \iteme{MSTP(92) :} (D=4) (C) energy partitioning in hadron or resolved photon remnant. The energy fraction $\chi$ taken by one of the two objects, with conventions as described for {\tt PARP(94) - PARP(98)}, is chosen according to the different distributions below. Here $c_{\mmin} = 2 \langle m_{\q} \rangle/E_{\mrm{cm}} = 0.6$ GeV$/E_{\mrm{cm}}$. \begin{subentry} \iteme{= 1 :} 1 for meson or resolved photon, $2(1-\chi)$ for baryon, i.e. simple counting rules. \iteme{= 2 :} $(k+1)(1-\chi)^k$, with $k$ as given in {\tt PARP(94) - PARP(97)}. \iteme{= 3 :} as {\tt =2} for remnant splitting into hadron plus jet, but proportional to $(1-\chi)^k/\sqrt[4]{\chi^2+c_{\mmin}^2}$ for remnant splitting into two jets, with $k$ given by {\tt PARP(94)} or {\tt PARP(96)}. \iteme{= 4 :} as {\tt =2} for remnant splitting into hadron plus jet, but proportional to $(1-\chi)^k/\sqrt{\chi^2+c_{\mmin}^2}$ for remnant splitting into two jets, with $k$ given by {\tt PARP(94)} or {\tt PARP(96)}. \iteme{= 5 :} as {\tt =2} for remnant splitting into hadron plus jet, but proportional to $(1-\chi)^k/(\chi^2+c_{\mmin}^2)^{b/2}$ for remnant splitting into two jets, with $k$ given by {\tt PARP(94)} or {\tt PARP(96)}, and $b$ by {\tt PARP(98)}. \end{subentry} \iteme{MSTP(93) :} (D=1) (C) primordial $k_{\perp}$ distribution in photon, either it is one of the incoming particles or inside an electron. \begin{subentry} \iteme{= 0 :} no primordial $k_{\perp}$. \iteme{= 1 :} Gaussian, width given in {\tt PARP(99)}, upper cut-off in {\tt PARP(100)}. \iteme{= 2 :} exponential, width given in {\tt PARP(99)}, upper cut-off in {\tt PARP(100)}. \iteme{= 3 :} power-like of the type $\d k_{\perp}^2/(k_{\perp 0}^2 + k_{\perp}^2)^2$, with $k_{\perp 0}$ in {\tt PARP(99)} and upper $k_{\perp}$ cut-off in {\tt PARP(100)}. \iteme{= 4 :} power-like of the type $\d k_{\perp}^2/(k_{\perp 0}^2 + k_{\perp}^2)$, with $k_{\perp 0}$ in {\tt PARP(99)} and upper $k_{\perp}$ cut-off in {\tt PARP(100)}. \iteme{= 5 :} power-like of the type $\d k_{\perp}^2/(k_{\perp 0}^2 + k_{\perp}^2)$, with $k_{\perp 0}$ in {\tt PARP(99)} and upper $k_{\perp}$ cut-off given by the $\pT$ of the hard process or by {\tt PARP(100)}, whichever is smaller. \itemc{Note:} for options 1 and 2 the {\tt PARP(100)} value is of minor importance, once {\tt PARP(100)}$\gg${\tt PARP(99)}. However, options 3 and 4 correspond to distributions with infinite $\langle k_{\perp}^2 \rangle$ if the $k_{\perp}$ spectrum is not cut off, and therefore the {\tt PARP(100)} value is as important for the overall distribution as is {\tt PARP(99)}. \end{subentry} \iteme{MSTP(131) :}\label{p:MSTP131} (D=0) master switch for pile-up events, i.e. several independent hadron--hadron interactions generated in the same bunch--bunch crossing, with the events following one after the other in the event record. \begin{subentry} \iteme{= 0 :} off, i.e. only one event is generated at a time. \iteme{= 1 :} on, i.e. several events are allowed in the same event record. Information on the processes generated may be found in {\tt MSTI(41) - MSTI(50)}. \end{subentry} \iteme{MSTP(132) :} (D=4) the processes that are switched on for pile-up events. The first event may be set up completely arbitrarily, using the switches in the {\tt PYSUBS} common block, while all the subsequent events have to be of one of the `inclusive' processes which dominate the cross section, according to the options below. It is thus not possible to generate two rare events in the pile-up option. \begin{subentry} \iteme{= 1 :} low-$\pT$ processes (ISUB = 95) only. The low-$\pT$ model actually used, both in the hard event and in the pile-up events, is the one set by {\tt MSTP(81)} etc. This means that implicitly also high-$\pT$ jets can be generated in the pile-up events. \iteme{= 2 :} low-$\pT$ + double diffractive processes (ISUB = 95 and 94). \iteme{= 3 :} low-$\pT$ + double diffractive + single diffractive processes (ISUB = 95, 94, 93 and 92). \iteme{= 4 :} low-$\pT$ + double diffractive + single diffractive + elastic processes, together corresponding to the full hadron--hadron cross section (ISUB = 95, 94, 93, 92 and 91). \end{subentry} \iteme{MSTP(133) :} (D=0) multiplicity distribution of pile-up events. \begin{subentry} \iteme{= 0 :} selected by user, before each {\tt PYEVNT} call, by giving the {\tt MSTP(134)} value. \iteme{= 1 :} a Poissonian multiplicity distribution in the total number of pile-up events. This is the relevant distribution if the switches set for the first event in {\tt PYSUBS} give the same subprocesses as are implied by {\tt MSTP(132)}. In that case the mean number of events per beam crossing is $\br{n} = \sigma_{\mrm{pile}} \times${\tt PARP(31)}, where $\sigma_{\mrm{pile}}$ is the sum of the cross section for allowed processes. Since bunch crossing which do not give any events at all (probability $\exp(-\br{n})$) are not simulated, the actual average number per {\tt PYEVNT} call is $\langle n \rangle = \br{n}/(1-\exp(-\br{n}))$. \iteme{= 2 :} a biased distribution, as is relevant when one of the events to be generated is assumed to belong to an event class with a cross section much smaller than the total hadronic cross section. If $\sigma_{\mrm{rare}}$ is the cross section for this rare process (or the sum of the cross sections of several rare processes) and $\sigma_{\mrm{pile}}$ the cross section for the processes allowed by {\tt MSTP(132)}, then define $\br{n} = \sigma_{\mrm{pile}} \times${\tt PARP(131)} and $f = \sigma_{\mrm{rare}}/\sigma_{\mrm{pile}}$. The probability that a bunch crossing will give $i$ events is then ${\cal P}_i = f \, i \, \exp(-\br{n}) \, \br{n}^i/i!$, i.e. the na\"{\i}ve Poissonian is suppressed by a factor $f$ since one of the events will be rare rather than frequent, but enhanced by a factor $i$ since any of the $i$ events may be the rare one. Only beam crossings which give at least one event of the required rare type are simulated, and the distribution above normalized accordingly. \itemc{Note:} for practical reasons, it is required that $\br{n} < 120$, i.e. that an average beam crossing does not contain more than 120 pile-up events. The multiplicity distribution is truncated above 200, or when the probability for a multiplicity has fallen below $10^{-6}$, whichever occurs sooner. Also low multiplicities with probabilities below $10^{-6}$ are truncated. See also {\tt PARI(91) - PARI(93)}. \end{subentry} \iteme{MSTP(134) :} (D=1) a user selected multiplicity, i.e. total number of pile-up events, to be generated in the next {\tt PYEVNT} call. May be reset for each new event, but must be in the range $1 \leq${\tt MSTP(134)}$\leq 200$. \boxsep \iteme{PARP(81) :}\label{p:PARP81} (D=1.40 GeV/$c$) effective minimum transverse momentum $\pTmin$ for multiple interactions with {\tt MSTP(82)=1}. \iteme{PARP(82) :} (D=1.55 GeV/$c$) regularization scale $p_{\perp 0}$ of the transverse momentum spectrum for multiple interactions with {\tt MSTP(82)}$\geq 2$. \iteme{PARP(83), PARP(84) :} (D=0.5, 0.2) parameters of an assumed double Gaussian matter distribution inside the colliding hadrons for {\tt MSTP(82)=4}, of the form given in eq. (\ref{mi:doubleGauss}), i.e. with a core of radius {\tt PARP(84)} of the main radius and containing a fraction {\tt PARP(83)} of the total hadronic matter. \iteme{PARP(85) :} (D=0.33) probability that an additional interaction in the multiple interaction formalism gives two gluons, with colour connections to `nearest neighbours' in momentum space. \iteme{PARP(86) :} (D=0.66) probability that an additional interaction in the multiple interaction formalism gives two gluons, either as described in {\tt PARP(85)} or as a closed gluon loop. Remaining fraction is supposed to consist of quark--antiquark pairs. \iteme{PARP(87), PARP(88) :} (D=0.7, 0.5) in order to account for an assumed dominance of valence quarks at low transverse momentum scales, a probability is introduced that a $\g\g$-scattering according to na\"{\i}ve cross section is replaced by a $\q\qbar$ one; this is used only for {\tt MSTP(82)}$\geq 2$. The probability is parametrized as ${\cal P} = a (1 - (\pT^2/(\pT^2 + b^2)^2)$, where $a =${\tt PARP(87)} and $b =${\tt PARP(88)}$\times${\tt PARP(82)}. \iteme{PARP(91) :} (D=0.44 GeV/$c$) (C) width of Gaussian primordial $k_{\perp}$ distribution inside hadron for {\tt MSTP(91)=1}, i.e. $\exp(-k_{\perp}^2/\sigma^2) \, k_{\perp} \, \d k_{\perp}$ with $\sigma =${\tt PARP(91)} and $\langle k_{\perp}^2 \rangle = ${\tt PARP(91)}$^2$. \iteme{PARP(92) :} (D=0.20 GeV/$c$) (C) width parameter of exponential primordial $k_{\perp}$ distribution inside hadron for {\tt MSTP(91)=2}, i.e. $\exp(-k_{\perp}/\sigma) \, k_{\perp} \, \d k_{\perp}$ with $\sigma =${\tt PARP(92)} and $\langle k_{\perp}^2 \rangle = 6 \times${\tt PARP(92)}$^2$. Thus one should put {\tt PARP(92)}$\approx${\tt PARP(91)}$/\sqrt{6}$ to have continuity with the option above. \iteme{PARP(93) :} (D=2. GeV/$c$) (C) upper cut-off for primordial $k_{\perp}$ distribution inside hadron. \iteme{PARP(94) :} (D=1.) (C) for {\tt MSTP(92)}$\geq 2$ this gives the value of the parameter $k$ for the case when a meson or resolved photon remnant is split into two fragments (which is which is chosen at random). \iteme{PARP(95) :} (D=0.) (C) for {\tt MSTP(92)}$\geq 2$ this gives the value of the parameter $k$ for the case when a meson or resolved photon remnant is split into a meson and a spectator fragment jet, with $\chi$ giving the energy fraction taken by the meson. \iteme{PARP(96) :} (D=3.) (C) for {\tt MSTP(92)}$\geq 2$ this gives the value of the parameter $k$ for the case when a nucleon remnant is split into a diquark and a quark fragment, with $\chi$ giving the energy fraction taken by the quark jet. \iteme{PARP(97) :} (D=1.) (C) for {\tt MSTP(92)}$\geq 2$ this gives the value of the parameter $k$ for the case when a nucleon remnant is split into a baryon and a quark jet or a meson and a diquark jet, with $\chi$ giving the energy fraction taken by the quark jet or meson, respectively. \iteme{PARP(98) :} (D=0.75) (C) for {\tt MSTP(92)=5} this gives the power of an assumed basic $1/\chi^b$ behaviour in the splitting distribution, with $b =${\tt PARP(98)}. \iteme{PARP(99) :} (D=0.44 GeV/$c$) (C) width parameter of primordial $k_{\perp}$ distribution inside photon; exact meaning depends on {\tt MSTP(93)} value chosen (cf. {\tt PARP(91)} and {\tt PARP(92)} above). \iteme{PARP(100) :} (D=2. GeV/$c$) (C) upper cut-off for primordial $k_{\perp}$ distribution inside photon. \iteme{PARP(131) :}\label{p:PARP131} (D=0.01 mb$^{-1}$) in the pile-up events scenario, {\tt PARP(131)} gives the assumed luminosity per bunch--bunch crossing, i.e. if a subprocess has a cross section $\sigma$, the average number of events of this type per bunch--bunch crossing is $\br{n} = \sigma \times${\tt PARP(131)}. {\tt PARP(131)} may be obtained by dividing the integrated luminosity over a given time (1 s, say) by the number of bunch--bunch crossings that this corresponds to. Since the program will not generate more than 200 pile-up events, the initialization procedure will crash if $\br{n}$ is above 120. \end{entry} \clearpage \section{Fragmentation} The main fragmentation option in {\JePy} is the Lund string scheme, but independent fragmentation options are also available. These latter options should not be taken too seriously, since we know that independent fragmentation does not provide a consistent alternative, but occasionally one may like to compare string fragmentation with something else. The subsequent four subsections give further details; the first one on flavour selection, which is common to the two approaches, the second on string fragmentation, the third on independent fragmentation, while the fourth and final contains information on a few other minor issues. The Lund fragmentation model is described in \cite{And83}, where all the basic ideas are presented and earlier papers \cite{And79,And80,And82,And82a} summarized. The details given there on how a multiparton jet system is allowed to fragment are out of date, however, and for this one should turn to \cite{Sjo84}. Also the `popcorn' baryon production mechanism is not covered, see \cite{And85}. Reviews of fragmentation models in general may be found in \cite{Sjo88,Sjo89}. \subsection{Flavour Selection} \label{ss:flavoursel} In either string or independent fragmentation, an iterative approach is used to describe the fragmentation process. Given an initial quark $\q = \q_0$, it is assumed that a new $\q_1 \qbar_1$ pair may be created, such that a meson $\q_0 \qbar_1$ is formed, and a $\q_1$ is left behind. This $\q_1$ may at a later stage pair off with a $\qbar_2$, and so on. What need be given is thus the relative probabilities to produce the various possible $\q_i \qbar_i$ pairs, $\u \ubar$, $\d \dbar$, $\s \sbar$, etc., and the relative probilities that a given $\q_{i-1}\qbar_i$ quark pair combination forms a specific meson, e.g. for $\u \ubar$ either $\pi^+$, $\rho^+$ or some higher state. In {\Je}, it is assumed that the two aspects can be factorized, i.e. that it is possible first to select a $\q_i \qbar_i$ pair, without any reference to allowed physical meson states, and that, once the $\q_{i-1} \qbar_i$ flavour combination is given, it can be assigned to a given meson state with total probability unity. \subsubsection{Quark flavours and transverse momenta} In order to generate the quark--antiquark pairs $\q_i \qbar_i$ which lead to string breakups, the Lund model invokes the idea of quantum mechanical tunnelling, as follows. If the $\q_i$ and $\qbar_i$ have no (common) mass or transverse momentum, the pair can classically be created at one point and then be pulled apart by the field. If the quarks have mass and/or transverse momentum, however, the $\q_i$ and $\qbar_i$ must classically be produced at a certain distance so that the field energy between them can be transformed into the sum of the two transverse masses $m_{\perp}$. Quantum mechanically, the quarks may be created in one point (so as to keep the concept of local flavour conservation) and then tunnel out to the classically allowed region. In terms of a common transverse mass $m_{\perp}$ of the $\q_i$ and the $\qbar_i$, the tunnelling probability is given by \begin{equation} \exp \left( -\frac{\pi m_{\perp}^2}{\kappa} \right) = \exp \left( -\frac{\pi m^2}{\kappa} \right) \exp \left( -\frac{\pi \pT^2}{\kappa} \right) ~. \end{equation} The factorization of the transverse momentum and the mass terms leads to a flavour-independent Gaussian spectrum for the $p_x$ and $p_y$ components of $\q_i \qbar_i$ pairs. Since the string is assumed to have no transverse excitations, this $\pT$ is locally compensated between the quark and the antiquark of the pair. The $\pT$ of a meson $\q_{i-1} \qbar_i$ is given by the vector sum of the $\pT$:s of the $\q_{i-1}$ and $\qbar_i$ constituents, which implies Gaussians in $p_x$ and $p_y$ with a width $\sqrt{2}$ that of the quarks themselves. The assumption of a Gaussian shape may be a good first approximation, but there remains the possibility of non-Gaussian tails, that can be important in some situations. In a perturbative QCD framework, a hard scattering is associated with gluon radiation, and further contributions to what is na\"{\i}vely called fragmentation $\pT$ comes from unresolved radiation. This is used as an explanation why the experimental $\left\langle \pT \right\rangle$ is somewhat higher than obtained with the formula above. The formula also implies a suppression of heavy quark production $u : d : s : c \approx$ \mbox{$1 : 1 : 0.3 : 10^{-11}$}. Charm and heavier quarks are hence not expected to be produced in the soft fragmentation. Since the predicted flavour suppressions are in terms of quark masses, which are notoriously difficult to assign (should it be current algebra, or constituent, or maybe something in between?), the suppression of $\s \sbar$ production is left as a free parameter in the program: $\u \ubar$ : $\d \dbar$ : $\s \sbar$ = 1 : 1 : $\gamma_s$, where by default $\gamma_s = 0.3$. At least qualitatively, the experimental value agrees with theoretical prejudice. There is no production at all of heavier flavours in the fragmentation process, but only as part of the shower evolution. \subsubsection{Meson production} \label{sss:mesonprod} Once the flavours $\q_{i-1}$ and $\qbar_i$ have been selected, a choice is made between the possible multiplets. The relative composition of different multiplets is not given from first principles, but must depend on the details of the fragmentation process. To some approximation one would expect a negligible fraction of states with radial excitations or non-vanishing orbital angular momentum. Spin counting arguments would then suggest a 3:1 mixture between the lowest lying vector and pseudoscalar multiplets. Wave function overlap arguments lead to a relative enhancement of the lighter pseudoscalar states, which is more pronounced the larger the mass splitting is \cite{And82a}. In the program, six meson multiplets are included. If the nonrelativistic classification scheme is used, i.e. mesons are assigned a valence quark spin $S$ and an internal orbital angular momentum $L$, with the physical spin $s$ denoted $J$, $\mbf{J} = \mbf{L} + \mbf{S}$, then the multiplets are: \begin{Itemize} \item $L = 0$, $S = 0$, $J = 0$: the ordinary pseudoscalar meson multiplet; \item $L = 0$, $S = 1$, $J = 1$: the ordinary vector meson multiplet; \item $L = 1$, $S = 0$, $J = 1$: an axial vector meson multiplet; \item $L = 1$, $S = 1$, $J = 0$: the scalar meson multiplet; \item $L = 1$, $S = 1$, $J = 1$: another axial vector meson multiplet; and \item $L = 1$, $S = 1$, $J = 2$: the tensor meson multiplet. \end{Itemize} Each multiplet has the full four-generation setup of $8 \times 8$ states included in the program, although many could never actually be produced. Some simplifications have been made; thus there is no mixing included between the two axial vector multiplets. In the program, the spin $S$ is first chosen to be either 0 or 1. This is done according to parametrized relative probabilities, where the probability for spin 1 by default is taken to be 0.5 for a meson consisting only of $\u$ and $\d$ quark, 0.6 for one which contains $\s$ as well, and $0.75$ for quarks with $\c$ or heavier quark, in accordance with the deliberations above. By default, it is assumed that $L = 0$, such that only pseudoscalar and vector mesons are produced. For inclusion of $L = 1$ production, four parameters can be used, one to give the probability that a $S = 0$ state also has $L =1$, the other three for the probability that a $S = 1$ state has $L = 1$ and $J$ either 0, 1, or 2. For the flavour-diagonal meson states $\u \ubar$, $\d \dbar$ and $\s \sbar$, it is also necessary to include mixing into the physical mesons. This is done according to a parametrization, based on the mixing angles given in the Review of Particle Properties \cite{PDG88}. In particular, the default choices correspond to \begin{eqnarray} \eta & = & \frac{1}{2} (\u\ubar + \d\dbar) - \frac{1}{\sqrt{2}} \s\sbar ~; \nonumber \\ \eta' & = & \frac{1}{2} (\u\ubar + \d\dbar) + \frac{1}{\sqrt{2}} \s\sbar ~; \nonumber \\ \omega & = & \frac{1}{\sqrt{2}} (\u\ubar + \d\dbar) \nonumber \\ \phi & = & \s\sbar ~. \end{eqnarray} In the $\pi^0 - \eta - \eta'$ system, no account is therefore taken of the difference in masses, an approximation which seems to lead to an overestimate of $\eta'$ rates \cite{ALE92}. Recently, parameters have been introduced to allow an additional `brute force' suppression of $\eta$ and $\eta'$ states. \subsubsection{Baryon production} Baryon production may, in its simplest form, be obtained by assuming that any flavour $\q_i$ given above could represent either a quark or an antidiquark in a colour triplet state. Then the same basic machinery can be run through as above, supplemented with the probability to produce various diquark pairs. In principle, there is one parameter for each diquark, but if tunnelling is still assumed to give an effective description, mass relations can be used to reduce the effective number of parameters. There are three main ones appearing in the program: \begin{Itemize} \item the relative probability to pick a $\qbar\qbar$ diquark rather than a $\q$; \item the extra suppression associated with a diquark containing a strange quark (over and above the ordinary $\s / \u$ suppression factor $\gamma_s$); and \item the suppression of spin 1 diquarks relative to spin 0 ones (apart from the factor of 3 enhancement of the former based on counting the number of spin states). \end{Itemize} The extra strange diquark suppression factor comes about since what appears in the exponent of the tunnelling formula is $m^2$ and not $m$, so that the diquark and the strange quark suppressions do not factorize. Only two baryon multiplets are included, i.e. there are no $L=1$ excited states. The two multiplets are: \begin{Itemize} \item $S = J = 1/2$: the `octet' multiplet of SU(3) (in the full four-generation scenario in the program 168 states are available); \item $S = J = 3/2$: the `decuplet' multiplet of SU(3) (120 states in the program). \end{Itemize} In contrast to the meson case, different flavour combinations have different numbers of states available: for $\u \u \u$ only $\Delta^{++}$, whereas $\u \d \s$ may become either $\Lambda$, $\Sigma^0$ or $\Sigma^{*0}$. An important constraint is that a baryon is a symmetric state of three quarks, neglecting the colour degree of freedom. When a diquark and a quark are joined to form a baryon, the combination is therefore weighted with the probability that they form a symmetric three-quark state. The program implementation of this principle is to first select a diquark at random, with the strangeness and spin 1 suppression factors above included, but then to accept the selected diquark with a weight proportional to the number of states available for the quark-diquark combination. This means that, were it not for the tunnelling suppression factors, all states in the {\bf SU(6)} (flavour {\bf SU(3)} times spin {\bf SU(2)}) 56-multiplet would become equally populated. Of course also heavier baryons may come from the fragmentation of e.g. $\c$ quark jets, but although the particle classification scheme used in the program is {\bf SU(16)}, i.e. with eight flavours, all possible quark-diquark combinations can be related to {\bf SU(6)} by symmetry arguments. As in the case for mesons, one could imagine an explicit further suppression of the heavier spin 3/2 baryons. We do not expect it to be an important effect, since baryon mass splittings are much smaller than in the meson case. In case of rejection, a new diquark is selected and tested, etc. A corresponding procedure is used for the quark selection when a diquark has already been formed in the previous step. Properly speaking both the quark and the diquark flavour should be chosen anew. This would become a tedious process, since also the hadron produced in the step before would have to be rejected. In practice only the last produced pair, be that the quark or diquark one, is rejected. The error introduced by this is small. A more general framework for baryon production is the `popcorn' one \cite{And85}, in which diquarks as such are never produced, but rather baryons appear from the successive production of several $\q_i \qbar_i$ pairs. The picture is the following. Assume that the original $\q$ is red $r$ and the $\qbar$ is $\br{r}$. Normally a new $\q_1 \qbar_1$ pair produced in the field would also be $r \br{r}$, so that the $\qbar_1$ is pulled towards the $\q$ end and vice versa, and two separate colour-singlet systems $\q \qbar_1$ and $\q_1 \qbar$ are formed. Occasionally, the $\q_1 \qbar_1$ pair may be e.g. $g \br{g}$ ($g$ = green), in which case there is no net colour charge acting on either $\q_1$ or $\qbar_1$. Therefore, the pair cannot gain energy from the field, and normally would exist only as a fluctuation. If $\q_1$ moves towards $\q$ and $\qbar_1$ towards $\qbar$, the net field remaining between $\q_1$ and $\qbar_1$ is $\br{b} b$ ($b$ = blue; $g + r = \br{b}$ if only colour triplets are assumed). In this central field, an additional $\q_2 \qbar_2$ pair can be created, where $\q_2$ now is pulled towards $\q \q_1$ and $\qbar_2$ towards $\qbar \qbar_1$, with no net colour field between $\q_2$ and $\qbar_2$. If this is all that happens, the baryon $B$ will be made up out of $\q_1$, $\q_2$ and some $\q_4$ produced between $\q$ and $\q_1$, and $\br{B}$ of $\qbar_1$, $\qbar_2$ and some $\qbar_5$, i.e. the $B$ and $\br{B}$ will be nearest neighbours in rank and share two quark pairs. Specifically, $\q_1$ will gain energy from $\q_2$ in order to end up on mass shell, and the tunnelling formula for an effective $\q_1 \q_2$ diquark is recovered. Part of the time, several $b \br{b}$ colour pair productions may take place between the $\q_1$ and $\qbar_1$, however. With two production vertices $\q_2 \qbar_2$ and $\q_3 \qbar_3$, a central meson $\qbar_2 \q_3$ may be formed, surrounded by a baryon $\q_4 \q_1 \q_2$ and an antibaryon $\qbar_3 \qbar_1 \qbar_5$. We call this a $BM\br{B}$ configuration to distinguish it from the $\q_4 \q_1 \q_2$ + $\qbar_2 \qbar_1 \qbar_5$ $B\br{B}$ configuration above. For $BM\br{B}$ the $B$ and $\br{B}$ only share one quark--antiquark pair, as opposed to two for $B\br{B}$ configurations. The relative probability for a $BM\br{B}$ configuration is given by the uncertainty relation suppression for having the $\q_1$ and $\qbar_1$ sufficiently far apart that a meson may be formed in between. Strictly speaking, also configurations like $BMM\br{B}$, $BMMM\br{B}$, etc. should be possible, but the probability for this is small in our model. Further, since larger masses corresponds to longer string pieces, the production of pseudoscalar mesons is favoured over that of vector ones. If only $B\br{B}$ and $BM\br{B}$ states are included, and if the probability for having a vector meson $M$ is not suppressed extra, two partly compensating errors are made (since a vector meson typically decays into two or more pseudoscalar ones). In total, the flavour iteration procedure therefore contains the following possible subprocesses (plus, of course, their charge conjugates): \begin{Itemize} \item $\q_1 \to \q_2 + (\q_1 \qbar_2)$ meson; \item $\q_1 \to \qbar_2\qbar_3 + (\q_1 \q_2 \q_3)$ baryon; \item $\q_1 \q_2 \to \qbar_3 + (\q_1 \q_2 \q_3)$ baryon; \item $\q_1 \q_2 \to \q_1 \q_3 + (\q_2 \qbar_3)$ meson; \end{Itemize} with the constraint that the last process cannot be iterated to obtain several mesons in between the baryon and the antibaryon. Unfortunately, the resulting baryon production model has a fair number of parameters, which would be given by the model only if quark and diquark masses were known unambiguously. We have already mentioned the $\s / \u$ ratio and the $\q\q / \q$ one; the latter has to be increased from 0.09 to 0.10 for the popcorn model, since the total number of possible baryon production configurations is lower in this case (the particle produced between the $B$ and $\br{B}$ is constrained to be a meson). For the popcorn model, exactly the same parameters as already found in the diquark model are needed to describe the $B\br{B}$ configurations. For $BM\br{B}$ configurations, the square root of a suppression factor should be applied if the factor is relevant only for one of the $B$ and $\br{B}$, e.g. if the $B$ is formed with a spin 1 `diquark' $\q_1 \q_2$ but the $\br{B}$ with a spin 0 diquark $\qbar_1 \qbar_3$. Additional parameters include the relative probability for $BM\br{B}$ configurations, which is assumed to be roughly 0.5 (with the remaining 0.5 being $B\br{B}$), a suppression factor for having a strange meson $M$ between the $B$ and $\br{B}$ (as opposed to having a lighter nonstrange one) and a suppression factor for having a $\s \sbar$ pair (rather than a $\u \ubar$ one) shared between the $B$ and $\br{B}$ of a $BM\br{B}$ configuration. The default parameter values are based on a combination of experimental observation and internal model predictions. In the diquark model, a diquark is expected to have exactly the same transverse momentum distribution as a quark. For $BM\br{B}$ configurations the situation is somewhat more unclear, but we have checked that various possibilities give very similar results. The option implemented in the program is to assume no transverse momentum at all for the $\q_1 \qbar_1$ pair shared by the $B$ and $\br{B}$, with all other pairs having the standard Gaussian spectrum with local momentum conservation. This means that the $B$ and $\br{B}$ $\pT$:s are uncorrelated in a $BM\br{B}$ configuration and (partially) anticorrelated in the $B\br{B}$ configurations, with the same mean transverse momentum for primary baryons as for primary mesons. Occasionally, the endpoint of a string is not a single parton, but a diquark or antidiquark, e.g. when a quark has been kicked out of a proton beam particle. One could consider fairly complex schemes for the resulting fragmentation. One such \cite{And81} was available in {\Je} version 6 but is no longer found in version 7. Instead the same basic scheme is used as for diquark pair production above. Thus a $\q\q$ diquark endpoint is let to fragment just as would a $\q\q$ produced in the field behind a matching $\qbar\qbar$ flavour, i.e. either the two quarks of the diquark enter into the same leading baryon, or else a meson is first produced, containing one of the quarks, while the other is contained in the baryon produced in the next step. \subsection{String Fragmentation} An iterative procedure can also be used for other aspects of the fragmentation. This is possible because, in the string picture, the various points where the string break by the production of $\q \qbar$ pairs are causally disconnected. Whereas the space--time picture in the c.m. frame is such that slow particles (in the middle of the system) are formed first, this ordering is Lorentz frame dependent and hence irrelevant. One may therefore make the convenient choice of starting an iteration process at the ends of the string and proceeding towards the middle. The string fragmentation scheme is rather complicated for a generic multiparton state. In order to simplify the discussion, we will therefore start with the simple $\q \qbar$ process, and only later survey the complications that appear when additional gluons are present. (This distinction is made for pedagogical reasons, in the program there is only one general-purpose algorithm). \subsubsection{Fragmentation functions} Assume a $\q \qbar$ jet system, in its c.m. frame, with the quark moving out in the $+z$ direction and the antiquark in the $-z$ one. We have discussed how it is possible to start the flavour iteration from the $\q$ end, i.e. pick a $\q_1 \qbar_1$ pair, form a hadron $\q \qbar_1$, etc. It has also been noted that the tunnelling mechanism is assumed to give a transverse momentum $\pT$ for each new $\q_i\qbar_i$ pair created, with the $\pT$ locally compensated between the $\q_i$ and the $\qbar_i$ member of the pair, and with a Gaussian distribution in $p_x$ and $p_y$ separately. In the program, this is regulated by one parameter, which gives the root-mean-square $\pT$ of a quark. Hadron transverse momenta are obtained as the sum of $\pT$:s of the constituent $\q_i$ and $\qbar_{i+1}$, where a diquark is considered just as a single quark. What remains to be determined is the energy and longitudinal momentum of the hadron. In fact, only one variable can be selected independently, since the momentum of the hadron is constrained by the already determined hadron transverse mass $m_{\perp}$, \begin{equation} (E+p_z)(E-p_z) = E^2 - p_z^2 = m_{\perp}^2 = m^2 + p_x^2 + p_y^2 ~. \label{fr:massconstr} \end{equation} In an iteration from the quark end, one is led (by the desire for longitudinal boost invariance and other considerations) to select the $z$ variable as the fraction of $E+p_z$ taken by the hadron, out of the available $E+p_z$. As hadrons are split off, the $E+p_z$ (and $E-p_z$) left for subsequent steps is reduced accordingly: \begin{eqnarray} (E+p_z)_{\mrm{new}} & = & (1-z) (E+p_z)_{\mrm{old}} ~, \nonumber \\ (E-p_z)_{\mrm{new}} & = & (E-p_z)_{\mrm{old}} - \frac{m_{\perp}^2}{z (E+p_z)_{\mrm{old}}} ~. \end{eqnarray} The fragmentation function $f(z)$, which expresses the probability that a given $z$ is picked, could in principle be arbitrary --- indeed, several such choices can be used inside the program, see below. If one, in addition, requires that the fragmentation process as a whole should look the same, irrespectively of whether the iterative procedure is performed from the $\q$ end or the $\qbar$ one, `left--right symmetry', the choice is essentially unique \cite{And83a}: the `Lund symmetric fragmentation function', \begin{equation} f(z) \propto \frac{1}{z} z^{a_{\alpha}} \left( \frac{1-z}{z} \right)^{a_{\beta}} \exp \left( - \frac{bm_{\perp}^2}{z} \right) ~. \label{fr:LSFFlong} \end{equation} There is one separate parameter $a$ for each flavour, with the index $\alpha$ corresponding to the `old' flavour in the iteration process, and $\beta$ to the `new' flavour. It is customary to put all $a_{\alpha,\beta}$ the same, and thus arrive at the simplified expression \begin{equation} f(z) \propto z^{-1} (1-z)^a \exp (-bm_{\perp}^2/z) ~. \label{fr:LSFF} \end{equation} In the program, only two separate $a$ values can be given, that for quark pair production and that for diquark one; by default the two are taken to be the same. In addition, there is the $b$ parameter, which is universal. It should be noted that the explicit mass dependence in $f(z)$ implies a harder fragmentation function for heavier hadrons; the asymptotic behaviour of the mean $z$ value for heavy hadrons is \begin{equation} \langle z \rangle \approx 1 - \frac{1+a}{bm_{\perp}^2} ~. \end{equation} Unfortunately it seems this predicts a somewhat harder spectrum for $\B$ mesons than observed in data. For future reference we note that the derivation of $f(z)$ as a by-product also gives the probability distribution in invariant time $\tau$ of $\q_i \qbar_i$ breakup vertices. In terms of $\Gamma = (\kappa \tau)^2$, this distribution is \begin{equation} {\cal P}(\Gamma) \, \d \Gamma \propto \Gamma^a \, \exp(-b \Gamma) \, \d \Gamma ~, \end{equation} with the same $a$ and $b$ as above. Many different other fragmentation functions have been proposed, and a few are available as options in the program. \begin{Itemize} \item The Field-Feynman parametrization \cite{Fie78}, \begin{equation} f(z) = 1 - a + 3a(1-z)^2 ~, \end{equation} with default value $a = 0.77$, is intended to be used only for ordinary hadrons made out of $\u$, $\d$ and $\s$ quarks. \item Since there are indications that the shape above is too strongly peaked at $z = 0$, instead a shape like \begin{equation} f(z) = (1+c) (1-z)^c \end{equation} may be used. \item Charm and bottom data clearly indicate the need for a harder fragmentation function for heavy flavours. The best known of these is the Peterson et al.~formula \cite{Pet83} \begin{equation} f(z) \propto \frac{1}{ z \left( 1 - \frac{\displaystyle 1}{\displaystyle z} - \frac{\displaystyle \epsilon_Q}{\displaystyle 1-z} \right)^2 } ~, \label{fr:PetHF} \end{equation} where $\epsilon_Q$ is a free parameter, expected to scale between flavours like $\epsilon_Q \propto 1/m_Q^2$. \item As a crude alternative, that is also peaked at $z=1$, one may use \begin{equation} f(z) = (1+c) z^c ~. \end{equation} \item Bowler \cite{Bow81} has shown, within the framework of the Artru--Mennessier model \cite{Art74}, that a massive endpoint quark with mass $m_Q$ leads to a modification of the symmetric fragmentation function, due to the fact that the string area swept out is reduced for massive endpoint quarks, compared with massless ditto. The Artru--Mennessier model in principle only applies for clusters with a continuous mass spectrum, and does not allow an $a$ term (i.e. $a \equiv 0$); however, it has been shown \cite{Mor89} that, for a discrete mass spectrum, one may still retain an effective $a$ term. In the program an approximate form with an $a$ term has therefore been used: \begin{equation} f(z) \propto \frac{1}{z^{1 + r_Q b m_Q^2}} z^{a_{\alpha}} \left( \frac{1-z}{z} \right)^{a_{\beta}} \exp \left( - \frac{b m_{\perp}^2}{z} \right) ~. \label{fr:LSFFBowler} \end{equation} In principle the prediction is that $r_Q \equiv 1$, but so as to be able to extrapolate smoothly between this form and the Lund symmetric one, it is possible to pick $r_Q$ separately for $\c$, $\b$ and $\t$ hadrons. \end{Itemize} \subsubsection{Joining the jets} The $f(z)$ formula above is only valid, for the breakup of a jet system into a hadron plus a remainder-system, when the remainder mass is large. If the fragmentation algorithm were to be used all the way from the $\q$ end to the $\qbar$ one, the mass of the last hadron to be formed at the $\qbar$ end would be completely constrained by energy and momentum conservation, and could not be on its mass shell. In theory it is known how to take such effects into account, but the resulting formulae are wholly unsuitable for Monte Carlo implementation. The practical solution to this problem is to carry out the fragmentation both from the $\q$ and the $\qbar$ end, such that for each new step in the fragmentation process, a random choice is made as to from what side the step is to be taken. If the step is on the $\q$ side, then $z$ is interpreted as fraction of the remaining $E+p_z$ of the system, while $z$ is interpreted as $E-p_z$ fraction for a step from the $\qbar$ end. At some point, when the remaining mass of the system has dropped below a given value, it is decided that the next breakup will produce two final hadrons, rather than a hadron and a remainder-system. Since the momenta of two hadrons are to be selected, rather than that of one only, there are enough degrees of freedom to have both total energy and total momentum completely conserved. The mass at which the normal fragmentation process is stopped and the final two hadrons formed is not actually a free parameter of the model: it is given by the requirement that the string everywhere looks the same, i.e. that the rapidity spacing of the final two hadrons, internally and with respect to surrounding hadrons, is the same as elsewhere in the fragmentation process. The stopping mass, for a given setup of fragmentation parameters, has therefore been determined in separate runs. If the fragmentation parameters are changed, some retuning should be done but, in practice, reasonable changes can be made without any special arrangements. Consider a fragmentation process which has already split off a number of hadrons from the $\q$ and $\qbar$ sides, leaving behind a a $\q_i \qbar_j$ remainder system. When this system breaks by the production of a $\q_n \qbar_n$ pair, it is decided to make this pair the final one, and produce the last two hadrons $\q_i\qbar_n$ and $\q_n\qbar_j$, if \begin{equation} ( (E+p_z)(E-p_z) )_{\mrm{remaining}} = W_{\mrm{rem}}^2 < W_{\mmin}^2 ~. \end{equation} The $W_{\mmin}$ is calculated according to \begin{equation} W_{\mmin} = ( W_{\mmin 0} + m_{\q i} + m_{\q j} + k \, m_{\q n} ) \, (1 \pm \delta) ~. \end{equation} Here $W_{\mmin 0}$ is the main free parameter, typically around 1 GeV, determined to give a flat rapidity plateau (separately for each particle species), while the default $k = 2$ corresponds to the mass of the final pair being taken fully into account. Smaller values may also be considered, depending on what criteria are used to define the `best' joining of the $\q$ and the $\qbar$ chain. The factor $1 \pm \delta$, by default evenly distributed between 0.8 and 1.2, signifies a smearing of the $W_{\mmin}$ value, to avoid an abrupt and unphysical cut-off in the invariant mass distribution of the final two hadrons. Still, this distribution will be somewhat different from that of any two adjacent hadrons elsewhere. Due to the cut there will be no tail up to very high masses; there are also fewer events close to the lower limit, where the two hadrons are formed at rest with respect to each other. This procedure does not work all that well for heavy flavours, since it does not fully take into account the harder fragmentation function encountered. Therefore, in addition to the check above, one further test is performed for charm and heavier flavours, as follows. If the check above allows more particle production, a heavy hadron $\q_i \qbar_n$ is formed, leaving a remainder $\q_n \qbar_j$. The range of allowed $z$ values, i.e. the fraction of remaining $E+p_z$ that may be taken by the $\q_i \qbar_n$ hadron, is constrained away from 0 and 1 by the $\q_i \qbar_n$ mass and minimal mass of the $\q_n \qbar_j$ system. The limits of the physical $z$ range is obtained when the $\q_n \qbar_j$ system only consists of one single particle, which then has a well-determined transverse mass $m_{\perp}^{(0)}$. From the $z$ value obtained with the infinite-energy fragmentation function formulae, a rescaled $z'$ value between these limits is given by \begin{equation} z' = \frac{1}{2} \left\{ 1 + \frac{m_{\perp i n}^2}{W_{\mrm{rem}}^2} - \frac{m_{\perp n j}^{(0)2}}{W_{\mrm{rem}}^2} + \sqrt{ \left( 1 - \frac{m_{\perp i n}^2}{W_{\mrm{rem}}^2} - \frac{m_{\perp n j}^{(0)2}}{W_{\mrm{rem}}^2} \right)^2 - 4 \frac{m_{\perp i n}^2}{W_{\mrm{rem}}^2} \frac{m_{\perp n j}^{(0)2}}{W_{\mrm{rem}}^2} } \, (2 z -1) \right\} ~. \end{equation} From the $z'$ value, the actual transverse mass $m_{\perp n j} \geq m_{\perp n j}^{(0)}$ of the $\q_n \qbar_j$ system may be calculated. For more than one particle to be produced out of this system, the requirement \begin{equation} m_{\perp n j}^2 = (1-z') \, \left( W_{\mrm{rem}}^2 - \frac{m_{\perp i n}^2}{z'} \right) > (m_{q j} + W_{\mmin 0})^2 + \pT^2 \end{equation} has to be fulfilled. If not, the $\q_n \qbar_j$ system is assumed to collapse to one single particle. The consequence of the procedure above is that, the more the infinite energy fragmentation function $f(z)$ is peaked close to $z=1$, the more likely it is that only two particles are produced. In particular, for $\t \tbar$ systems, where very large $\langle z \rangle$ values are predicted, the expectation is that two particle final states will dominate far above the threshold region. The procedure above has been constructed so that the two particle fraction can be calculated directly from the shape of $f(z)$ and the (approximate) mass spectrum, but it is not unique. For the symmetric Lund fragmentation function, a number of alternatives tried all give essentially the same result, whereas other fragmentation functions may be more sensitive to details. Assume now that two final hadrons have been picked. If the transverse mass of the remainder-system is smaller than the sum of transverse masses of the final two hadrons, the whole fragmentation chain is rejected, and started over from the $\q$ and $\qbar$ endpoints. This does not introduce any significant bias, since the decision to reject a fragmentation chain only depends on what happens in the very last step, specifically that the next-to-last step took away too much energy, and not on what happened in the steps before that. If, on the other hand, the remainder-mass is large enough, there are two kinematically allowed solutions for the final two hadrons: the two mirror images in the rest frame of the remainder-system. Also the choice between these two solutions is given by the consistency requirements, and can be derived from studies of infinite energy jets. The probability for the reverse ordering, i.e. where the rapidity and the flavour orderings disagree, is parametrized by \begin{equation} {\cal P}_{\mrm{reverse}} = \frac{1}{2} \left( \frac{m_{\perp i n} + m_{\perp n j}}{W_{\mrm{rem}}} \right)^d ~. \label{fr:revord} \end{equation} For symmetric fragmentation, the ordering is expected to be increasingly strict when the particles involved are more massive. In the program it is therefore assumed that $d$ is a function of the masses, $d = d_0 (m_{\perp i n} + m_{\perp n j})^2$, where $d_0$ is a free parameter. When baryon production is included, some particular problems arise. First consider $B\br{B}$ situations. In the na\"{\i}ve iterative scheme, away from the middle of the event, one already has a quark and is to chose a matching diquark flavour or the other way around. In either case the choice of the new flavour can be done taking into account the number of {\bf SU(6)} states available for the quark-diquark combination. For a case where the final $\q_n \qbar_n$ breakup is an antidiquark-diquark one, the weights for forming $\q_i \qbar_n$ and $\q_n \qbar_i$ enter at the same time, however. We do not know how to handle this problem; what is done is to use weights as usual for the $\q_i \qbar_n$ baryon to select $\q_n$, but then consider $\q_n \qbar_i$ as given (or the other way around with equal probability). If $\q_n \qbar_i$ turns out to be an antidiquark-diquark combination, the whole fragmentation chain is rejected, since we do not know how to form corresponding hadrons. A similar problem arises, and is solved in the same spirit, for a $BM\br{B}$ configuration in which the $B$ (or $\br{B}$) was chosen as third-last particle. When only two particles remain to be generated, it is obviously too late to consider having a $BM\br{B}$ configuration. This is as it should, however, as can be found by looking at all possible ways a hadron of given rank can be a baryon. While some practical compromises have to be accepted in the joining procedure, the fact that the joining takes place in different parts of the string in different events means that, in the end, essentially no visible effects remain. \subsubsection{String motion and infrared stability} We have now discussed the SF scheme for the fragmentation of a simple $\q \qbar$ jet system. In order to understand how these results generalize to arbitrary jet systems, it is first necessary to understand the string motion for the case when no fragmentation takes place. In the following we will assume that quarks as well as gluons are massless, but all arguments can be generalized to massive quarks without too much problem. For a $\q \qbar$ event viewed in the c.m. frame, with total energy $W$, the partons start moving out back-to-back, carrying half the energy each. As they move apart, energy and momentum is lost to the string. When the partons are a distance $W / \kappa$ apart, all the energy is stored in the string. The partons now turn around and come together again with the original momentum vectors reversed. This corresponds to half a period of the full string motion; the second half the process is repeated, mirror-imaged. For further generalizations to multiparton systems, a convenient description of the energy and momentum flow is given in terms of `genes' \cite{Art83}, infinitesimal packets of the four-momentum given up by the partons to the string. Genes with $p_z = E$, emitted from the $\q$ end in the initial stages of the string motion above, will move in the $\qbar$ direction with the speed of light, whereas genes with $p_z = -E$ given up by the $\qbar$ will move in the $\q$ direction. Thus, in this simple case, the direction of motion for a gene is just opposite to that of a free particle with the same four-momentum. This is due to the string tension. If the system is not viewed in the c.m. frame, the rules are that any parton gives up genes with four-momentum proportional to its own four-momentum, but the direction of motion of any gene is given by the momentum direction of the genes it meets, i.e. that were emitted by the parton at the other end of that particular string piece. When the $\q$ has lost all its energy, the $\qbar$ genes, which before could not catch up with $\q$, start impinging on it, and the $\q$ is pulled back, accreting $\qbar$ genes in the process. When the $\q$ and $\qbar$ meet in the origin again, they have completely traded genes with respect to the initial situation. A 3-jet $\q \qbar \g$ event initially corresponds to having a string piece stretched between $\q$ and $\g$ and another between $\g$ and $\qbar$. Gluon four-momentum genes are thus flowing towards the $\q$ and $\qbar$. Correspondingly, $\q$ and $\qbar$ genes are flowing towards the $\g$. When the gluon has lost all its energy, the $\g$ genes continue moving apart, and instead a third string region is formed in the `middle' of the total string, consisting of overlapping $\q$ and $\qbar$ genes. The two `corners' on the string, separating the three string regions, are not of the gluon-kink type: they do not carry any momentum. If this third region would only appear at a time later than the typical time scale for fragmentation, it could not affect the sharing of energy between different particles. This is true in the limit of high energy, well separated partons. For a small gluon energy, on the other hand, the third string region appears early, and the overall drawing of the string becomes fairly 2-jet-like, since the third string region consists of $\q$ and $\qbar$ genes and therefore behaves exactly as a sting pulled out directly between the $\q$ and $\qbar$. In the limit of vanishing gluon energy, the two initial string regions collapse to naught, and the ordinary 2-jet event is recovered. Also for a collinear gluon, i.e. $\theta_{\q \g}$ (or $\theta_{\qbar \g}$) small, the stretching becomes 2-jet-like. In particular, the $\q$ string endpoint first moves out a distance $\mbf{p}_{\q} / \kappa$ losing genes to the string, and then a further distance $\mbf{p}_{\g} / \kappa$, a first half accreting genes from the $\g$ and the second half re-emitting them. (This latter half actually includes yet another string piece; a corresponding piece appears at the $\qbar$ end, such that half a period of the system involves five different string regions.) The end result is, approximately, that a string is drawn out as if there had only been a single parton with energy $|\mbf{p}_{\q} + \mbf{p}_{\g}|$, such that the simple 2-jet event again is recovered in the limit $\theta_{\q \g} \to 0$. These properties of the string motion are the reason why the string fragmentation scheme is `infrared safe' with respect to soft or collinear gluon emission. The discussions for the 3-jet case can be generalized to the motion of a string with $\q$ and $\qbar$ endpoints and an arbitrary number of intermediate gluons. For $n$ partons, whereof $n-2$ gluons, the original string contains $n-1$ pieces. Anytime one of the original gluons has lost its energy, a new string region is formed, delineated by a pair of `corners'. As the extra `corners' meet each other, old string regions vanish and new are created, so that half a period of the string contains $2n^2 - 6n + 5$ different string regions. Each of these regions can be understood simply as built up from the overlap of (opposite-moving) genes from two of the original partons, according to well specified rules. \subsubsection{Fragmentation of multiparton systems} The full machinery needed for a multiparton system is very complicated, and is described in detail in \cite{Sjo84}. The following outline is far from complete, and is complicated nonetheless. The main message to be conveyed is that a Lorentz covariant algorithm exists for handling an arbitrary parton configuration, but that the necessary machinery is more complex than in either cluster or independent fragmentation. Assume $n$ partons, with ordering along the string, and related four-momenta, given by $\q(p_1) \g(p_2) \g(p_3) \cdots \g(p_{n-1}) \qbar(p_n)$. The initial string then contains $n-1$ separate pieces. The string piece between the quark and its neigbouring gluon is, in four-momentum space, spanned by one side with four-momentum $p_+^{(1)} = p_1$ and another with $p_-^{(1)} = p_2/2$. The factor of 1/2 in the second expression comes from the fact that the gluon shares its energy between two string pieces. The indices `$+$' and `$-$' denotes direction towards the $\q$ and $\qbar$ end, respectively. The next string piece, counted from the quark end, is spanned by $p_+^{(2)} = p_2/2$ and $p_-^{(2)} = p_3/2$, and so on, with the last one being $p_+^{(n-1)} = p_{n-1}/2$ and $p_-^{(n-1)} = p_n$. For the algorithm to work, it is important that all $p_{\pm}^{(i)}$ be light-cone-like, i.e. $p_{\pm}^{(i)2} = 0$. Since gluons are massless, it is only the two endpoint quarks which can cause problems. The procedure here is to create new $p_{\pm}$ vectors for each of the two endpoint regions, defined to be linear combinations of the old $p_{\pm}$ ones for the same region, with coefficients determined so that the new vectors are light-cone-like. De facto, this corresponds to replacing a massive quark at the end of a string piece with a massless quark at the end of a somewhat longer string piece. With the exception of the added fictitious piece, which anyway ends up entirely within the heavy hadron produced from the heavy quark, the string motion remains unchanged by this. In the continued string motion, when new string regions appear as time goes by, the first such string regions that appear can be represented as being spanned by one $p_+^{(j)}$ and another $p_-^{(k)}$ four-vector, with $j$ and $k$ not necessarily adjacent. For instance, in the $\q \g \qbar$ case, the `third' string region is spanned by $p_+^{(1)}$ and $p_-^{(3)}$. Later on in the string evolution history, it is also possible to have regions made up of two $p_+$ or two $p_-$ momenta. These appear when an endpoint quark has lost all its original momentum, has accreted the momentum of an gluon, and is now re-emitting this momentum. In practice, these regions may be neglected. Therefore only pieces made up by a $(p_+^{(j)},p_-^{(k)})$ pair of momenta are considered in the program. The allowes string regions may be ordered in an abstract parameter plane, where the $(j,k)$ indices of the four-momentum pairs define the position of each region along the two (parameter plane) coordinate axes. In this plane the fragmentation procedure can be described as a sequence of steps, starting at the quark end, where $(j,k) = (1,1)$, and ending at the antiquark one, $(j,k) = (n-1,n-1)$. Each step is taken from an `old' $\q_{i-1} \qbar_{i-1}$ pair production vertex, to the production vertex of a `new' $\q_i \qbar_i$ pair, and the string piece between these two string breaks represent a hadron. Some steps may be taken within one and the same region, while others may have one vertex in one region and the other vertex in another region. Consistency requirements, like energy-momentum conservation, dictates that vertex $j$ and $k$ region values be ordered in a monotonic sequence, and that the vertex positions are monotonically ordered inside each region. The four-momentum of each hadron can be read off, for $p_+$ ($p_-$) momenta, by projecting the separation between the old and the new vertex on to the $j$ ($k$) axis. If the four-momentum fraction of $p_{\pm}^{(i)}$ taken by a hadron is denoted $x_{\pm}^{(i)}$, then the total hadron four-momentum is given by \begin{equation} p = \sum_{j=j_1}^{j_2} x_+^{(j)} p_+^{(j)} + \sum_{k=k_1}^{k_2} x_-^{(k)} p_-^{(k)} + p_{x1} \hat{e}_x^{(j_1 k_1)} + p_{y1} \hat{e}_y^{(j_1 k_1)} + p_{x2} \hat{e}_x^{(j_2 k_2)} + p_{y2} \hat{e}_y^{(j_2 k_2)} ~, \label{fr:fourmom} \end{equation} for a step from region $(j_1,k_1)$ to region $(j_2,k_2)$. By necessity, $x_+^{(j)}$ is unity for a $j_1 < j < j_2$, and correspondingly for $x_-^{(k)}$. The $(p_x,p_y)$ pairs are the transverse momenta produced at the two string breaks, and the $(\hat{e}_x,\hat{e}_y)$ pairs four-vectors transverse to the string directions in the regions of the respective string breaks: \begin{eqnarray} & & \hat{e}_x^{(jk)2} = \hat{e}_y^{(jk)2} = -1 ~, \nonumber \\ & & \hat{e}_x^{(jk)} \hat{e}_y^{(jk)} = \hat{e}_{x,y}^{(jk)} p_+^{(j)} = \hat{e}_{x,y}^{(jk)} p_-^{(k)} = 0 ~. \end{eqnarray} The fact that the hadron should be on mass shell, $p^2 = m^2$, puts one constraint on where a new breakup may be, given that the old one is already known, just as eq. (\ref{fr:massconstr}) did in the simple 2-jet case. The remaining degree of freedom is, as before, to be given by the fragmentation function $f(z)$. The interpretation of the $z$ is only well-defined for a step entirely constrained to one of the initial string regions, however, which is not enough. In the 2-jet case, the $z$ values can be related to the invariant times of string breaks, as follows. The variable $\Gamma = (\kappa \tau)^2$, with $\kappa$ the string tension and $\tau$ the invariant time between the production vertex of the partons and the breakup point, obeys an iterative relation of the kind \begin{eqnarray} \Gamma_0 & = & 0 ~, \nonumber \\ \Gamma_i & = & (1-z_i) \left( \Gamma_{i-1} + \frac{m_{\perp i}^2}{z_i} \right) ~. \end{eqnarray} Here $\Gamma_0$ represents the value at the $\q$ and $\qbar$ endpoints, and $\Gamma_{i-1}$ and $\Gamma_i$ the values at the old and new breakup vertices needed to produce a hadron with transverse mass $m_{\perp i}$, and with the $z_i$ of the step chosen according to $f(z_i)$. The invariant time can be defined in an unambiguous way, also over boundaries between the different string regions, so for multijet events the $z$ variable may be interpreted just as an auxiliary variable needed to determine the next $\Gamma$ value. (In the Lund symmetric fragmentation function derivation, the $\Gamma$ variable actually does appear naturally, so the choice is not as arbitrary as it may seem here.) The mass and $\Gamma$ constraints together are sufficient to determine where the next string breakup is to be chosen, given the preceding one in the iteration scheme. Actually, several ambiguities remain, but are of no importance for the overall picture. The algorithm for finding the next breakup then works something like follows. Pick a hadron, $\pT$, and $z$, and calculate the next $\Gamma$. If the old breakup is in the region $(j,k)$, and if the new breakup is also assumed to be in the same region, then the $m^2$ and $\Gamma$ constraints can be reformulated in terms of the fractions $x_+^{(j)}$ and $x_-^{(k)}$ the hadron must take of the total four-vectors $p_+^{(j)}$ and $p_-^{(k)}$: \begin{eqnarray} m^2 &=& c_1 + c_2 x_+^{(j)} + c_3 x_-^{(k)} + c_4 x_+^{(j)} x_-^{(k)} ~, \nonumber \\ \Gamma &=& d_1 + d_2 x_+^{(j)} + d_3 x_-^{(k)} + d_4 x_+^{(j)} x_-^{(k)} ~. \label{fr:mGa} \end{eqnarray} Here the coefficients $c_n$ are fairly simple expressions, obtainable by squaring eq. (\ref{fr:fourmom}), while $d_n$ are slightly more complicated in that they depend on the position of the old string break, but both the $c_n$ and the $d_n$ are explicitly calculable. What remains is an equation system with two unknowns, $x_+^{(j)}$ and $x_-^{(k)}$. The absence of any quadratic terms is due to the fact that all $p_{\pm}^{(i)2} = 0$, i.e. to the choice of a formulation based on light-cone-like longitudinal vectors. Of the two possible solutions to the equation system (elimination of one variable gives a second degree equation in the other), one is unphysical and can be discarded outright. The other solution is checked for whether the $x_{\pm}$ values are actually inside the physically allowed region, i.e. whether the $x_{\pm}$ values of the current step, plus whatever has already been used up in previous steps, are less than unity. If yes, a solution has been found. If no, it is because the breakup could not take place inside the region studied, i.e. because the equation system was solved for the wrong region. One therefore has to change either index $j$ or index $k$ above by one step, i.e. go to the next nearest string region. In this new region, a new equation system of the type in eq. (\ref{fr:mGa}) may be written down, with new coefficients. A new solution is found and tested, and so on until a physically acceptable solution is found. The hadron four-momentum is now given by an expression of the type (\ref{fr:fourmom}). The breakup found forms the starting point for the new step in the fragmentation chain, and so on. The final joining in the middle is done as in the 2-jet case, with minor extensions. \subsection{Independent Fragmentation} The independent fragmentation (IF) approach dates back to the early seventies \cite{Krz72}, and gained widespread popularity with the Field-Feynman paper \cite{Fie78}. Subsequently, IF was the basis for two programs widely used in the early PETRA/PEP days, the Hoyer et al.~\cite{Hoy79} and the Ali et al.~\cite{Ali80} programs. JETSET has as (non-default) options a wide selection of independent fragmentation algorithms. \subsubsection{Fragmentation of a single jet} In the IF approach, it is assumed that the fragmentation of any system of partons can be described as an incoherent sum of independent fragmentation procedures for each parton separately. The process is to be carried out in the overall c.m. frame of the jet system, with each jet fragmentation axis given by the direction of motion of the corresponding parton in that frame. Exactly as in string fragmentation, an iterative ansatz can be used to describe the sucessive production of one hadron after the next. Assume that a quark is kicked out by some hard interaction, carrying a well-defined amount of energy and momentum. This quark jet $\q$ is split into a hadron $\q \qbar_1$ and a remainder-jet $\q_1$, essentially collinear with each other. New quark and hadron flavours are picked as already described. The sharing of energy and momentum is given by some probability distribution $f(z)$, where $z$ is the fraction taken by the hadron, leaving $1-z$ for the remainder-jet. The remainder-jet is assumed to be just a scaled-down version of the original jet, in an average sense. The process of splitting off a hadron can therefore be iterated, to yield a sequence of hadrons. In particular, the function $f(z)$ is assumed to be the same at each step, i.e. independent of remaining energy. If $z$ is interpreted as the fraction of the jet $E+p_{\mrm{L}}$, i.e. energy plus longitudinal momentum with respect to the jet axis, this leads to a flat central rapidity plateau $dn/dy$ for a large initial energy. Fragmentation functions can be chosen among those listed above for string fragmentation, but also here the default is the Lund symmetric fragmentation function. The normal $z$ interpretation means that a choice of a $z$ value close to $0$ corresponds to a particle moving backwards, i.e. with $p_{\mrm{L}} < 0$. It makes sense to allow only the production of particles with $p_{\mrm{L}} > 0$, but to explicitly constrain $z$ accordingly would destroy longitudinal invariance. The most straightforward way out is to allow all $z$ values but discard hadrons with $p_{\mrm{L}} < 0$. Flavour, transverse momentum and $E + p_{\mrm{L}}$ carried by these hadrons are `lost' for the forward jet. The average energy of the final jet comes out roughly right this way, with a spread of 1--2 GeV around the mean. The jet longitudinal momentum is decreased, however, since the jet acquires an effective mass during the fragmentation procedure. For a 2-jet event this is as it should be, at least on average, because also the momentum of the compensating opposite-side parton is decreased. In addition to local flavour conservation in $\q_i \qbar_i$ splittings, it is also assumed that transverse momentum is locally conserved, i.e. the net $\pT$ of the $\q_i \qbar_i$ pair as a whole is assumed to be vanishing. The $\pT$ of the $\q$ is taken to be a Gaussian in the two transverse degrees of freedom separately, with the transverse momentum of a hadron obtained by the sum of constituent quark transverse momenta. Within the IF framework, there is no unique recipe for how gluon jet fragmentation should be handled. One possibility is to treat it exactly like a quark jet, with the initial quark flavour chosen at random among $\u$, $\ubar$, $\d$, $\dbar$, $\s$ and $\sbar$, including the ordinary $\s$ quark suppression factor. Since the gluon is supposed to fragment more softly than a quark jet, the fragmentation fuction may be chosen independently. Another common option is to split the $\g$ jet into a pair of parallel $\q$ and $\qbar$ ones, sharing the energy, e.g. as in a perturbative branching $\g \to \q \qbar$, i.e. $f(z) \propto z^2 + (1-z)^2$. The fragmentation function could still be chosen independently, if so desired. Further, in either case the fragmentation $\pT$ could be chosen to have a different mean. \subsubsection{Fragmentation of a jet system} In a system of many jets, each jet is fragmented independently. Since each jet by itself does not conserves the flavour, energy and momentum, then neither does a system of jets. At the end of the generation, special algorithms are therefore used to patch this up. The choice of approach has major consequences, e.g. for event shapes and $\alphas$ determinations \cite{Sjo84a}. Little attention is usually given to flavour conservation, and we only offer one scheme. When the fragmentation of all jets has been performed, independently of each other, the net initial flavour composition, i.e. number of $\u$ quarks minus number of $\ubar$ quarks etc., is compared with the net final flavour composition. In case of an imbalance, the flavours of the hadron with lowest three-momentum are removed, and the imbalance is re-evaluated. If the remaining imbalance could be compensated by a suitable choice of new flavours for this hadron, flavours are so chosen, a new mass is found and the new energy can be evaluated, keeping the three-momentum of the original hadron. If the removal of flavours from the hadron with lowest momentum is not enough, flavours are removed from the one with next-lowest momentum, and so on until enough freedom is obtained, whereafter the necessary flavours are recombined at random to form the new hadrons. Occasionally one extra $\q_i \qbar_i$ pair must be created, which is then done according to the customary probabilities. Several different schemes for energy and momentum conservation have been devised. One \cite{Hoy79} is to conserve transverse momentum locally within each jet, so that the final momentum vector of a jet is always parallel with that of the corresponding parton. Then longitudinal momenta may be rescaled separately for particles within each jet, such that the ratio of rescaled jet momentum to initial parton momentum is the same in all jets. Since the initial partons had net vanishing three-momentum, so do now the hadrons. The rescaling factors may be chosen such that also energy comes out right. Another common approach \cite{Ali80} is to boost the event to the frame where the total hadronic momentum is vanishing. After that, energy conservation can be obtained by rescaling all particle three-momenta by a common factor. The number of possible schemes is infinite. Two further options are available in the program. One is to shift all particle three-momenta by a common amount to give net vanishing momentum, and then rescale as before. Another is to shift all particle three-momenta, for each particle by an amount proportional to the longitudinal mass with respect to the imbalance direction, and with overall magnitude selected to give momentum conservation, and then rescale as before. In addition, there is a choice of whether to treat separate colour singlets (like $\q \qbar'$ and $\q' \qbar$ in a $\q \qbar \q' \qbar'$ event) separately or as one single big system. A serious conceptual weakness of the IF framework is the issue of Lorentz invariance. The outcome of the fragmentation procedure depends on the coordinate frame chosen, a problem circumvented by requiring fragmentation always to be carried out in the c.m. frame. This is a consistent procedure for 2-jet events, but only a technical trick for multijets. It should be noted, however, that a Lorentz covariant generalization of the independent fragmentation model exists, in which separate `gluon-type' and `quark-type' strings are used, the Montvay scheme \cite{Mon79}. The `quark string' is characterized by the ordinary string constant $\kappa$, whereas a `gluon string' is taken to have a string constant $\kappa_{\g}$. If $\kappa_{\g} > 2 \kappa$ it is always energetically favourable to split a gluon string into two quark ones, and the ordinary Lund string model is recovered. Otherwise, for a 3-jet $\q \qbar \g$ event the three different string pieces are joined at a junction. The motion of this junction is given by the composant of string tensions acting on it. In particular, it is always possible to boost an event to a frame where this junction is at rest. In this frame, much of the standard na\"{\i}ve IF picture holds for the fragmentation of the three jets; additionally, a correct treatment would automatically give flavour, momentum and energy conservation. Unfortunately, the simplicity is lost when studying events with several gluon jets. In general, each event will contain a number of different junctions, resulting in a polypod shape with a number of quark and gluons strings sticking out from a skeleton of gluon strings. In order to have a first look what this kind of scheme might mean for 3-jet events in $\ee$, the following kind of very simplified treatment was devised, using the framework already provided by the independent fragmentation routines. It is assumed that all strings of the system come together in one single junction. The frame where this junction is at rest may then be found iteratively. It turns out that the iteration scheme is unstable when having a parton with (almost) vanishing momentum in the rest frame of the junction. Since such a parton anyhow would not affect the motion of the junction for long, the effective string tension is scaled down linearly for partons with momenta smaller than 2 GeV. When the rest frame of the junction has been found, all partons are fragmented independently in that frame . The products are boosted back to the c.m. frame for the (optional) conservation of flavour, momentum and energy. A second conceptual weakness of IF is the issue of collinear divergences. In a parton-shower picture, where a quark or gluon is expected to branch into several reasonably collimated partons, the independent fragmentation of one single parton or of a bunch of collinear ones gives quite different outcomes, e.g. with a much larger hadron multiplicity in the latter case. It is conceivable that a different set of fragmentation functions could be constructed in the shower case in order to circumvent this problem (local parton--hadron duality \cite{Dok89} would correspond to having $f(z) = \delta(z-1)$). \subsection{Other Fragmentation Aspects} Here two aspects are considered, which are applicable regardless of whether string or independent fragmentation is used. \subsubsection{Small mass systems} Occasionally, a jet system may have too small an invariant mass for the ordinary jet fragmentation schemes. This is particularly a problem when showers are used, since two nearby $\g \to \q' \qbar'$ branchings may give rise to an intermediate low-mass colour-singlet system. Before the ordinary fragmentation, one includes an optional additional step, to catch situations of this kind. First the jet system with lowest invariant mass, minus endpoint quark masses, is found. If this is too low for jet fragmentation, an attempt is made to split the system into two hadrons by producing a new $\q_n \qbar_n$ pair (with $\q_n$ chosen according to the standard fragmentation scheme, so that e.g. also diquarks are allowed) to go with the existing endpoint flavours. If the sum of the two thus constructed hadron masses is smaller than the total invariant mass, a simple isotropic two-particle decay is performed. If not, the endpoint flavours are combined to give one single hadron. Next, the parton (or hadron) is found which, when taken together with the jet system, has the largest invariant mass. A minimal transfer of four-momentum is then performed, which puts the hadron on mass shell while keeping the mass of the parton unchanged. With this done, one may again search for a low-mass jet system, and iterate the procedure above, if need be. The procedure may be seen as a `poor man's cluster fragmentation', i.e. a cluster and a low-mass string are considered to be more or less the same thing. \subsubsection{Bose--Einstein effects} A crude option for the simulation of Bose--Einstein effects is included, but is turned off by default. Here the detailed physics is not that well understood, see e.g. \cite{Lor89}. What is offered is an algorithm, more than just a parametrization (since very specific assumptions and choices have been made), and yet less than a true model (since the underlying physics picture is rather fuzzy). In this scheme, the fragmentation is allowed to proceed as usual, and so is the decay of short-lived particles like $\rho$. Then pairs of identical particles, $\pi^+$ say, are considered one by one. The $Q_{ij}$ value of a pair $i$ and $j$ is evaluated, \begin{equation} Q_{ij} = \sqrt{ (p_i + p_j)^2 - 4m^2} ~, \end{equation} where $m$ is the common particle mass. A shifted (smaller) $Q'_{ij}$ is then to be found such that the (infinite statistics) ratio $C_2(Q)$ of shifted to unshifted $Q$ distributions is given by the requested parametrization. The shape may be chosen either exponential or Gaussian, \begin{equation} C_2(Q) = 1 + \lambda \exp \left( - (Q/d)^r \right), ~~~~r = 1~\mrm{or}~2 ~. \end{equation} (In fact, the distribution has to dip slightly below unity at $Q$ values outside the Bose enhancement region, from conservation of total multiplicity.) If the inclusive distribution of $Q_{ij}$ values is assumed given just by phase space, at least at small relative momentum then, with $\d^3 p / E \propto Q^2 \, \d Q / \sqrt{Q^2 + 4m^2}$, then $Q'_{ij}$ is found as the solution to the equation \begin{equation} \int_0^{Q_{ij}} \frac{Q^2 \, \d Q}{\sqrt{Q^2 + 4 m^2}} = \int_0^{Q'_{ij}} C_2(Q) \, \frac{Q^2 \, \d Q}{\sqrt{Q^2 + 4 m^2}} ~. \end{equation} The change of $Q_{ij}$ can be translated into an effective shift of the three-momenta of the two particles, if one uses as extra constraint that the total three-momentum of each pair be conserved in the c.m. frame of the event. Only after all pairwise momentum shifts have been evaluated, with respect to the original momenta, are these momenta actually shifted, for each particle by the sum of evaluated shifts. The total energy of the event is slightly reduced in the process, which is compensated by an overall rescaling of all c.m. frame momentum vectors. Finally, the decay chain is resumed with more long-lived particles like $\pi^0$. Two comments can be made. The Bose--Einstein effect is here interpreted almost as a classical force acting on the `final state', rather than as a quantum mechanical phenomenon on the production amplitude. This is not a credo, but just an ansatz to make things manageable. Also, since only pairwise interactions are considered, the effects associated with three or more nearby particles tend to get overestimated. (More exact, but also more time-consuming methods may be found in \cite{Zaj87}.) Thus the input $\lambda$ may have to be chosen smaller than what one wants to get out. (On the other hand, many of the pairs of an event contains at least one particle produced in some secondary vertex, like a $\D$ decay. This reduces the fraction of pairs which may contribute to the Bose--Einstein effects, and thus reduces the potential signal.) This option should therefore be used with caution, and only as a first approximation to what Bose--Einstein effects can mean. \clearpage \section{Particles and Their Decays} Particles are the building blocks from which events are constructed. We here use the word `particle' in its broadest sense, i.e. including partons, resonances, hadrons, and so on, subgroups we will describe in the following. Each particle is characterized by some quantities, such as charge and mass. In addition, many of the particles are unstable and subsequently decay. This section contains a survey of the particle content of the programs, and the particle properties assumed. In particular, the decay treatment is discussed. Some particle and decay properties form part already of the hard subprocess description, and are therefore described in sections \ref{s:JETSETproc}, \ref{s:PYTprocgen} and \ref{s:pytproc}. \subsection{The Particle Content} \label{ss:decpartcont} In order to describe both current and potential future physics, a number of different particles are needed. A list of some particles, along with their codes, is given in section \ref{ss:codes}. Here we therefore emphasize the generality rather than the details. Four full generations of quarks and leptons are included in the program, although indications from LEP strongly suggest that only three exist in Nature. There is no standard terminology for the fourth generation; we use $\lrm$ for the down type quark ($\lrm$ for low), $\hrm$ for the up type quark ($\hrm$ for high), $\chi$ for the lepton and $\nu_{\chi}$ for the neutrino. Quarks may appear either singly or in pairs; the latter are called diquarks and are characterized by their flavour content and their spin. A diquark is always assumed to be in a colour antitriplet state. From the coloured quarks (and diquarks), the colour neutral hadrons may be build up. Six full meson multiplets are included and two baryon ones, see section \ref{ss:flavoursel}. In addition, $\K_{\mrm{S}}^0$ and $\K_{\mrm{L}}^0$ are considered as separate particles coming from the `decay' of $\K^0$ and $\br{\K}^0$ (or, occasionally, produced directly). Other particles from the Standard Model include the gluon $\g$, the photon $\gamma$, the intermediate gauge bosons $\Z^0$ and $\W^{\pm}$, and the standard Higgs $\H^0$. Non-standard particles include additional gauge bosons, $\Z'^0$ and $\W'^{\pm}$, additional Higgs bosons $\H'^0$, $\A^0$ and $\H^{\pm}$, a leptoquark $\L_{\Q}$ and a horizontal gauge boson $\R^0$. It is also possible to use the particle codes of the current fourth generation fermions to represent excited quarks and leptons, $\q^*$ and $\ell^*$. From the point of view of usage inside the programs, particles may be subdivided into three classes, partly overlapping. \begin{Enumerate} \item A parton is generically any object which may be found in the wave function of the incoming beams, and may participate in initial- or final-state showers. This includes what is normally meant by partons, i.e. quarks and gluons, but here also leptons and photons. In a few cases other particles may be classified as partons in this sense. \item A resonance is an unstable particle produced as part of the hard process, and where the decay treatment normally is also part of the hard process. Resonance partial widths are perturbatively calculable, and therefore it is possible to dynamically recalculate branching ratios as a function of the mass assigned to a resonance. Resonances includes particles like the $\Z^0$ and other massive gauge bosons and Higgs particles. It does not include hadrons with non-vanishing width, like the $\rho$, which are just called `unstable hadrons'. \item Hadrons, i.e. mesons and baryons produced either in the fragmentation process, in secondary decays or as part of the beam remnant treatment, but not directly as part of the hard process (except in a few special cases). Hadrons may be stable or unstable. Branching ratios are not assumed perturbatively calculable, and can therefore be set freely. Also leptons and photons produced in decays belong to this class. \end{Enumerate} Usually the subdivision above is easy to understand and gives you the control you would expect. However, the classification of top and the fourth generation fermions may lead to some confusion, as already mentioned, section \ref{sss:heavflavclass}. The problem is that the top did not use to be treated as a resonance, but was rather allowed to fragment to hadrons, which subsequently decayed. This approach was a reasonable choice in the days when the top mass was assumed quite light by today's standards. However, given current top limits, the fragmentation and the decay of the top quark is being played out on comparable time scales, and the treatment becomes much more difficult (see e.g. ref. \cite{Sjo92a} for a toy model description). Starting at masses of around 120 GeV the top decay time is so short that no top hadrons at all are formed, and then a true resonance description is appropriate, but still with some complications due to the net colour charge of the top quark. Such an option is now default, wherein the top quark is assumed to decay immediately, but intermediate scenarios can not be modelled. The appearance of intermediate top hadrons in the description has little influence on event shapes, even for a very heavy top. \subsection{Masses, Widths and Lifetimes} \subsubsection{Masses} Quark masses are not particularly well defined. In the program it is necessary to make use of two kinds of masses, current algebra ones and constituent ones. The former are relevant for the kinematics in hard processes (e.g. in $\g\g \to \c\cbar$) and for couplings to Higgs particles, and therefore directly affect cross sections. These values are the ones stored in the standard mass array {\tt PMAS}. Constituent masses are used to derive the masses of hadrons, and are stored separately in the {\tt PARF} array. We maintain this distinction for the five first flavours, using the following values by default: \\ \begin{tabular}{ccc@{\protect\rule{0mm}{\tablinsep}}} quark & current algebra mass & constituent mass \\ d & 0.0099 GeV & 0.325 GeV \\ u & 0.0056 GeV & 0.325 GeV \\ s & 0.199 GeV & 0.5 GeV \\ c & 1.35 GeV & 1.6 GeV \\ b & 5.0 GeV & 5.0 GeV. \\ \end{tabular} \\ For top and fourth generation quarks the distinction is not as important, so only one set of mass values is used, namely the one in {\tt PMAS}. The default top mass is 160 GeV. Constituent masses for diquarks are defined as the sum of the respective quark masses. The gluon is always assumed massless. Particle masses, when known, are taken from ref. \cite{PDG92}. Hypothesized particles, such as fourth generation fermions and Higgs bosons, are assigned some not unreasonable set of default values, in the sense of where you want to search for them in the not too distant future. Here it is understood that you will go in and change the default values according to your own opinions at the beginning of a run. The total number of hadrons in the program is very large, whereof many are not yet discovered (like charm and bottom baryons) and other may or may not exist (top and fourth generation hadrons). In particular for the latter, it would be messy for the user to have to recalculate the masses of hadron each time the assumed quark mass is changed. Therefore the masses of yet undiscovered mesons and baryons are built up, when needed, from the constituent masses. For this purpose one uses formulae of the type \cite{DeR75} \begin{equation} m = m_0 + \sum_i m_i + k \, m_{\d}^2 \sum_{i 0$. \end{entry} \drawbox{CALL LU4ENT(IP,KF1,KF2,KF3,KF4,PECM,X1,X2,X4,X12,X14)}% \label{p:LU4ENT} \begin{entry} \itemc{Purpose:} to add four entries to the event record, i.e. either a 4-jet system or four separate particles (or, for $\q\qbar\q'\qbar'$ events, two 2-jet systems). \iteme{IP :} normally line number for the first jet/particle, with other three in lines {\tt IP+1}, {\tt IP+2} and {\tt IP+3}. There are two exceptions. \\ If {\tt IP=0}, lines 1, 2, 3 and 4 are used and {\tt LUEXEC} is called. \\ If {\tt IP<0}, lines {\tt -IP} through {\tt -IP+3} are used, with status code {\tt K(I,1)=3}, i.e. with special colour connection information, so that a parton shower can be generated by a {\tt LUSHOW} call, followed by a {\tt LUEXEC} call, if so desired (only relevant for jets). \iteme{KF1,KF2,KF3,KF4 :} flavour codes for the four jets/particles. \iteme{PECM :} ($=E_{\mrm{cm}}$) the total energy of the system. \iteme{X1,X2,X4 :} $x_i = 2E_i/E_{\mrm{cm}}$, i.e. twice the energy fraction taken by the $i$'th jet. Thus $x_3 = 2 - x_1 - x_2 - x_4$, and need not be given. \iteme{X12,X14 :} $x_{ij} = 2 p_i p_j/E_{\mrm{cm}}^2$, i.e. twice the four-vector product of the momenta for jets $i$ and $j$, properly normalized. With the masses known, other $x_{ij}$ may be constructed from the $x_i$ and $x_{ij}$ given. Note that not all combinations of $x_i$ and $x_{ij}$ are inside the physically allowed region. \itemc{Remark:} the system is given in the c.m. frame, with the first jet going out in the $+z$ direction and the fourth jet lying in the $xz$-plane with $p_x > 0$. The second jet will have $p_y > 0$ and $p_y < 0$ with equal probability, with the third jet balancing this $p_y$ (this corresponds to a random choice between the two possible stereoisomers). \end{entry} \drawbox{CALL LUJOIN(NJOIN,IJOIN)}\label{p:LUJOIN} \begin{entry} \itemc{Purpose:} to connect a number of previously defined partons into a string configuration. Initially the partons must be given with status codes {\tt K(I,1)=} 1, 2 or 3. Afterwards the partons all have status code 3, i.e. are given with full colour-flow information. Compared to the normal way of defining a parton system, the partons need therefore not appear in the same sequence in the event record as they are assumed to do along the string. It is also possible to call {\tt LUSHOW} for all or some of the entries making up the string formed by {\tt LUJOIN}. \iteme{NJOIN:} the number of entries that are to be joined by one string. \iteme{IJOIN:} an one-dimensional array, of size at least {\tt NJOIN}. The {\tt NJOIN} first numbers are the positions of the partons that are to be joined, given in the order the partons are assumed to appear along the string. If the system consists entirely of gluons, the string is closed by connecting back the last to the first entry. \itemc{Remarks:} only one string (i.e. one colour singlet) may be defined per call, but one is at liberty to use any number of {\tt LUJOIN} calls for a given event. The program will check that the parton configuration specified makes sense, and not take any action unless it does. Note, however, that an initially sensible parton configuration may become nonsensical, if only some of the partons are reconnected, while the others are left unchanged. \end{entry} \drawbox{CALL LUGIVE(CHIN)}\label{p:LUGIVE} \begin{entry} \itemc{Purpose:} to set the value of any variable residing in the commmonblocks {\tt LUJETS}, {\tt LUDAT1}, {\tt LUDAT2}, {\tt LUDAT3}, {\tt LUDAT4}, {\tt LUDATR}, {\tt PYSUBS}, {\tt PYPARS}, {\tt PYINT1}, {\tt PYINT2}, {\tt PYINT3}, {\tt PYINT4}, {\tt PYINT5}, {\tt PYINT6}, or {\tt PYINT7}. This is done in a more controlled fashion than by directly including the common blocks in the user program, in that array bounds are checked and the old and new values for the variable changed are written to the output for reference. \iteme{CHIN :} character expression of length at most 100 characters, with requests for variables to be changed, stored in the form \\ {\tt variable1=value1;variable2=value2;variable3=value3}\ldots~. \\ Note that an arbitrary number of instructions can be stored in one call if separated by semicolons, and that blanks may be included anyplace. The variable$_i$ may be any single variable in the {\JePy} common blocks, and the value$_i$ must be of the correct integer, real or character (without extra quotes) type. Array indices and values must be given explicitly, i.e. cannot be variables in their own right. The exception is that the first index can be preceded by a {\tt C}, signifying that the index should be translated from normal KF to compressed KC code with a {\tt LUCOMP} call; this is allowed for the {\tt KCHG}, {\tt PMAS}, {\tt MDCY} and {\tt CHAF} arrays. If a value$_i$ is omitted, i.e. with the construction {\tt variable=}, the current value is written to the output, but the variable itself is not changed. \itemc{Remark :} The checks on array bounds are hardwired into this routine. Therefore, if some user changes array dimensions and {\tt MSTU(3)}, {\tt MSTU(6)} and/or {\tt MSTU(7)}, as allowed by other considerations, these changes will not be known to {\tt LUGIVE}. Normally this should not be a problem, however. \end{entry} \subsection{The JETSET Physics Routines} The physics routines form the major part of {\Je}, but once the initial jet/particle configuration has been specified and default parameter values changed, if so desired, only a {\tt LUEXEC} call is necessary to simulate the whole fragmentation and decay chain. Therefore a normal user will not directly see any of the other routines in this section. Some of them could be called directly, but the danger of faulty usage is then non-negligible. \drawbox{CALL LUEXEC}\label{p:LUEXEC} \begin{entry} \itemc{Purpose:} to administrate the fragmentation and decay chain. {\tt LUEXEC} may be called several times, but only entries which have not yet been treated (more precisely, which have $1 \leq${\tt K(I,1)}$\leq 10$) can be affected by further calls. This may apply if more jets/particles have been added by the user, or if particles previously considered stable are now allowed to decay. The actions that will be taken during a {\tt LUEXEC} call can be tailored extensively via the {\tt LUDAT1}--{\tt LUDAT3} common blocks, in particular by setting the {\tt MSTJ} values suitably. \end{entry} \boxsep \begin{entry} \iteme{SUBROUTINE LUPREP(IP) :}\label{p:LUPREP} to rearrange parton shower end products (marked with {\tt K(I,1)=3}) sequentially along strings; also to (optionally) allow small jet systems to collapse into two particles or one only, in the latter case with energy and momentum to be shuffled elsewhere in the event; also to perform checks that e.g. flavours of colour-singlet systems make sense. \iteme{SUBROUTINE LUSTRF(IP) :}\label{p:LUSTRF} to generate the fragmentation of an arbitrary colour-singlet jet system according to the Lund string fragmentation model. In many respects, this routine is the very heart and soul of {\Je}. \iteme{SUBROUTINE LUINDF(IP) :}\label{p:LUINDF} to handle the fragmentation of a jet system according to independent fragmentation models, and implement energy, momentum and flavour conservation, if so desired. Also the fragmentation of a single jet, not belonging to a jet system, is considered here (this is of course physical nonsense, but may sometimes be convenient for specific tasks). \iteme{SUBROUTINE LUDECY(IP) :}\label{p:LUDECY} to perform a particle decay, according to known branching ratios or different kinds of models, depending on our level of knowledge. Various matrix elements are included for specific processes. \iteme{SUBROUTINE LUKFDI(KFL1,KFL2,KFL3,KF) :}\label{p:LUKFDI} to generate a new quark or diquark flavour and to combine it with an existing flavour to give a hadron. \begin{subentry} \iteme{KFL1:} incoming flavour. \iteme{KFL2:} extra incoming flavour, e.g. for formation of final particle, where the flavours are completely specified. Is normally 0. \iteme{KFL3:} newly created flavour; is 0 if {\tt KFL2} is non-zero. \iteme{KF:} produced hadron. Is 0 if something went wrong (e.g. inconsistent combination of incoming flavours). \end{subentry} \iteme{SUBROUTINE LUPTDI(KFL,PX,PY) :}\label{p:LUPTDI} to give transverse momentum, e.g. for a $\q\qbar$ pair created in the colour field, according to independent Gaussian distributions in $p_x$ and $p_y$. \iteme{SUBROUTINE LUZDIS(KFL1,KFL3,PR,Z) :}\label{p:LUZDIS} to generate the longitudinal scaling variable $z$ in jet fragmentation, either according to the Lund symmetric fragmentation function, or according to a choice of other shapes. \iteme{SUBROUTINE LUBOEI :}\label{p:LUBOEI} to include Bose--Einstein effects according to a simple para\-metri\-zation. By default, this routine is not called. If called from {\tt LUEXEC}, this is done after the decay of short-lived resonances, but before the decay of long-lived ones. This means the routine should never be called directly by you, nor would effects be correctly simulated if decays are switched off. See {\tt MSTJ(51) - MSTJ(52)} for switching on the routine. \iteme{FUNCTION ULMASS(KF) :}\label{p:ULMASS} to give the mass for a parton/particle. \iteme{SUBROUTINE LUNAME(KF,CHAU) :}\label{p:LUNAME} to give the parton/particle name (as a string of type {\tt CHARACTER CHAU*16}). \iteme{FUNCTION LUCHGE(KF) :}\label{p:LUCHGE} to give three times the charge for a parton/particle. \iteme{FUNCTION LUCOMP(KF) :}\label{p:LUCOMP} to give the compressed parton/particle code KC for a given KF code, as required to find entry into mass and decay data tables. Also checks whether the given KF code is actually an allowed one (i.e. known by the program), and returns 0 if not. Note that KF may be positive or negative, while the resulting KC code is never negative. \iteme{SUBROUTINE LUERRM(MERR,MESSAG) :}\label{p:LUERRM} to keep track of the number of errors and warnings encountered, write out information on them, and abort the program in case of too many errors. \iteme{FUNCTION ULANGL(X,Y) :}\label{p:ULANGL} to calculate the angle from the $x$ and $y$ coordinates. \iteme{SUBROUTINE LULOGO :}\label{p:LULOGO} to write a titlepage for the {\JePy} programs. Called by LULIST(0). \iteme{BLOCK DATA LUDATA :}\label{p:LUDATA} to give default values for variables in the {\tt LUDAT1}, {\tt LUDAT2}, {\tt LUDAT3}, {\tt LUDAT4} and {\tt LUDATR} common blocks. \end{entry} \subsection{Event Study and Data Listing Routines} After an {\tt LUEXEC} call, the event generated is stored in the {\tt LUJETS} common block, and whatever physical variable is desired may be constructed from this record. An event may be rotated, boosted or listed, and particle data may be listed or modified. Via the functions {\tt KLU} and {\tt PLU} the values of some frequently appearing variables may be obtained more easily. As described in section \ref{ss:evanrout}, also more detailed event shape analyses may be performed simply. \drawbox{CALL LUROBO(THE,PHI,BEX,BEY,BEZ)}\label{p:LUROBO} \begin{entry} \itemc{Purpose:} to perform rotations and Lorentz boosts (in that order, if both in the same call) of jet/particle momenta and vertex position variables. \iteme{THE, PHI :} standard polar coordinates $\theta, \varphi$, giving the rotated direction of a momentum vector initially along the $+z$ axis. \iteme{BEX, BEY, BEZ :} gives the direction and size \mbox{\boldmath $\beta$} of a Lorentz boost, such that a particle initially at rest will have $\mbf{p}/E = $\mbox{\boldmath $\beta$} afterwards. \itemc{Remark:} all entries 1 through {\tt N} are affected by the transformation, unless lower and upper bounds are explicitly given by {\tt MSTU(1)} and {\tt MSTU(2)}, or if status code {\tt K(I,1)}$\leq 0$. \end{entry} \drawbox{ENTRY LUDBRB(IMI,IMA,THE,PHI,DBEX,DBEY,DBEZ)}\label{p:LUDBRB} \begin{entry} \itemc{Purpose:} to perform rotations and Lorentz boosts (in that order, if both in the same call) of jet/particle momenta and vertex position variables, for a specific range of entries, and with the boost vector given in double precision. Is entry to {\tt LUROBO}, mainly intended for internal use. \iteme{IMI, IMA :} range of entries affected by transformation, {\tt IMI}$\leq${\tt I}$\leq${\tt IMA}. \iteme{THE, PHI :} standard polar coordinates $\theta, \varphi$, giving the rotated direction of a momentum vector initially along the $+z$ axis. \iteme{DBEX, DBEY, DBEZ :} gives the direction and size \mbox{\boldmath $\beta$} of a Lorentz boost, such that a particle initially at rest will have $\mbf{p}/E = $\mbox{\boldmath $\beta$} afterwards. Is to be given in double precision. \itemc{Remark:} all entries with status codes {\tt K(I,1)>0} in the requested range are affected by the transformation. \end{entry} \drawbox{CALL LUEDIT(MEDIT)}\label{p:LUEDIT} \begin{entry} \itemc{Purpose:} to exclude unstable or undetectable jets/particles from the event record. One may also use {\tt LUEDIT} to store spare copies of events (specifically initial parton configuration) that can be recalled to allow e.g. different fragmentation schemes to be run through with one and the same parton configuration. Finally, an event which has been analyzed with {\tt LUSPHE}, {\tt LUTHRU} or {\tt LUCLUS} (see section \ref{ss:evanrout}) may be rotated to align the event axis with the $z$ direction. \iteme{MEDIT :} tells which action is to be taken. \begin{subentry} \iteme{= 0 :} empty ({\tt K(I,1)=0}) and documentation ({\tt K(I,1)>20}) lines are removed. The jets/particles remaining are compressed in the beginning of the {\tt LUJETS} common block and the {\tt N} value is updated accordingly. The event history is lost, so that information stored in {\tt K(I,3)}, {\tt K(I,4)} and {\tt K(I,5)} is no longer relevant. \iteme{= 1 :} as {\tt =0}, but in addition all jets/particles that have fragmented/decayed ({\tt K(I,1)>10}) are removed. \iteme{= 2 :} as {\tt =1}, but also all neutrinos and unknown particles (i.e. compressed code KC$ = 0$) are removed. \iteme{= 3 :} as {\tt =2}, but also all uncharged, colour neutral particles are removed, leaving only charged, stable particles (and unfragmented partons, if fragmentation has not been performed). \iteme{= 5 :} as {\tt =0}, but also all partons which have branched or been rearranged in a parton shower and all particles which have decayed are removed, leaving only the fragmenting parton configuration and the final-state particles. \iteme{= 11 :} remove lines with {\tt K(I,1)<0}. Update event history information (in {\tt K(I,3) - K(I,5)}) to refer to remaining entries. \iteme{= 12 :} remove lines with {\tt K(I,1)=0}. Update event history information (in {\tt K(I,3) - K(I,5)}) to refer to remaining entries. \iteme{= 13 :} remove lines with {\tt K(I,1)}$ = 11$, 12 or 15, except for any line with {\tt K(I,2)=94}. Update event history information (in {\tt K(I,3) - K(I,5)}) to refer to remaining entries. In particular, try to trace origin of daughters, for which the mother is decayed, back to entries not deleted. \iteme{= 14 :} remove lines with {\tt K(I,1)}$ = 13$ or 14, and also any line with {\tt K(I,2)=94}. Update event history information (in {\tt K(I,3) - K(I,5)}) to refer to remaining entries. In particular, try to trace origin of rearranged jets back through the parton-shower history to the shower initiator. \iteme{= 15 :} remove lines with {\tt K(I,1)>20}. Update event history information (in {\tt K(I,3) - K(I,5)}) to refer to remaining entries. \iteme{= 16 :} try to reconstruct missing daughter pointers of decayed particles from the mother pointers of decay products. These missing pointers typically come from the need to use {\tt K(I,4)} and {\tt K(I,5)} also for colour flow information. \iteme{= 21 :} all partons/particles in current event record are stored (as a spare copy) in bottom of common block {\tt LUJETS} (is e.g. done to save original partons before calling {\tt LUEXEC}). \iteme{= 22 :} partons/particles stored in bottom of event record with {\tt =21} are placed in beginning of record again, overwriting previous information there (so that e.g. a different fragmentation scheme can be used on the same partons). Since the copy at bottom is unaffected, repeated calls with {\tt =22} can be made. \iteme{= 23 :} primary partons/particles in the beginning of event record are marked as not fragmented or decayed, and number of entries {\tt N} is updated accordingly. Is simpe substitute for {\tt =21} plus {\tt =22} when no fragmentation/decay products precede any of the original partons/particles. \iteme{= 31 :} rotate largest axis, determined by {\tt LUSPHE}, {\tt LUTHRU} or {\tt LUCLUS}, to sit along the $z$ direction, and the second largest axis into the $xz$ plane. For {\tt LUCLUS} it can be further specified to $+z$ axis and $xz$ plane with $x > 0$, respectively. Requires that one of these routines has been called before. \iteme{= 32 :} mainly intended for {\tt LUSPHE} and {\tt LUTHRU}, this gives a further alignment of the event, in addition to the one implied by {\tt =31}. The `slim' jet, defined as the side ($z > 0$ or $z < 0$) with the smallest summed $\pT$ over square root of number of particles, is rotated into the $+z$ hemisphere. In the opposite hemisphere (now $z < 0$), the side of $x > 0$ and $x < 0$ which has the largest summed $|p_z|$ is rotated into the $z < 0, x > 0$ quadrant. Requires that {\tt LUSPHE} or {\tt LUTHRU} has been called before. \end{subentry} \itemc{Remark:} all entries 1 through {\tt N} are affected by the editing. For options 0--5 lower and upper bounds can be explicitly given by {\tt MSTU(1)} and {\tt MSTU(2)}. \end{entry} \drawbox{CALL LULIST(MLIST)}\label{p:LULIST} \begin{entry} \itemc{Purpose:} to list an event, jet or particle data, or current parameter values. \iteme{MLIST :} determines what is to be listed. \begin{subentry} \iteme{= 0 :} writes a title page, common for {\Je} and {\Py}, with program version numbers and last dates of change; is mostly for internal use. \iteme{= 1 :} gives a simple list of current event record, in an 80 column format suitable for viewing directly on the computer terminal. For each entry, the following information is given: the entry number {\tt I}, the parton/particle name (see below), the status code ({\tt K(I,1)}), the flavour code KF ({\tt K(I,2)}), the line number of the mother ({\tt K(I,3)}), and the three-momentum, energy and mass ({\tt P(I,1) - P(I,5)}). If {\tt MSTU(3)} is non-zero, lines immediately after the event record proper are also listed. A final line contains information on total charge, momentum, energy and invariant mass. \\ The particle name is given by a call to the routine {\tt LUNAME}. For an entry which has decayed/fragmented ({\tt K(I,1)=} 11--20), this particle name is given within parentheses. Similarly, a documentation line ({\tt K(I,1)=} 21--30) has the name enclosed in expression signs (!\ldots!) and an event/jet axis information line the name within inequality signs ($<$\ldots$>$). If the last character of the name is a `?', it is a signal that the complete name has been truncated to fit in, and can therefore not be trusted; this is very rare. For partons which have been arranged along strings ({\tt K(I,1)=} 1, 2, 11 or 12), the end of the parton name column contains information about the colour string arrangement: an {\tt A} for the first entry of a string, an {\tt I} for all intermediate ones, and a {\tt V} for the final one (a poor man's rendering of a vertical doublesided arrow, $\updownarrow$). \\ It is possible to insert lines just consisting of sequences of {\tt ======} to separate different sections of the event record, see {\tt MSTU(70) - MSTU(80)}. \iteme{= 2 :} gives a more extensive list of the current event record, in a 132 column format, suitable for printers or workstations. For each entry, the following information is given: the entry number {\tt I}, the parton/particle name (with padding as described for {\tt =1}), the status code ({\tt K(I,1)}), the flavour code KF ({\tt K(I,2)}), the line number of the mother ({\tt K(I,3)}), the decay product/colour-flow pointers ({\tt K(I,4), K(I,5)}), and the three-momentum, energy and mass ({\tt P(I,1) - P(I,5)}). If {\tt MSTU(3)} is non-zero, lines immediately after the event record proper are also listed. A final line contains information on total charge, momentum, energy and invariant mass. Lines with only {\tt ======} may be inserted as for {\tt =1}. \iteme{= 3 :} gives the same basic listing as {\tt =2}, but with an additional line for each entry containing information on production vertex position and time ({\tt V(I,1) - V(I,4)}) and, for unstable particles, invariant lifetime ({\tt V(I,5)}). \iteme{= 11 :} provides a simple list of all parton/particle codes defined in the program, with KF code and corresponding particle name. The list is grouped by particle kind, and only within each group in ascending order. \iteme{= 12 :} provides a list of all parton/particle and decay data used in the program. Each parton/particle code is represented by one line containing KF flavour code, KC compressed code, particle name, antiparticle name (where appropriate), electrical and colour charge (stored in {\tt KCHG}), mass, resonance width and maximum broadening, average invariant lifetime (in {\tt PMAS}) and whether the particle is considered stable or not (in {\tt MDCY}). Immediately after a particle, each decay channel gets one line, containing decay channel number ({\tt IDC} read from {\tt MDCY}), on/off switch for the channel, matrix element type ({\tt MDME}), branching ratio ({\tt BRAT}), and decay products ({\tt KFDP}). The {\tt MSTU(14)} flag can be used to set the maximum flavour for which particles are listed, with the default (= 0) corresponding to separately defined ones (KC$ > 100$ if KF$ > 0$). In order to keep the size down, decay modes of heavy hadrons collectively defined are never listed; these have KC codes 84--88, where the relevant information may be found. \iteme{= 13 :} gives a list of current parameter values for {\tt MSTU}, {\tt PARU}, {\tt MSTJ} and {\tt PARJ}, and the first 200 entries of {\tt PARF}. This is useful to keep check of which default values were changed in a given run. \end{subentry} \itemc{Remark:} for options 1--3 and 12 lower and upper bounds of the listing can be explicitly given by {\tt MSTU(1)} and {\tt MSTU(2)}. \end{entry} \drawbox{CALL LUUPDA(MUPDA,LFN)}\label{p:LUUPDA} \begin{entry} \itemc{Purpose:} to give you the ability to update particle data, or to keep several versions of modified particle data for special purposes (e.g. charm studies). \iteme{MUPDA :} gives the type of action to be taken. \begin{subentry} \iteme{= 1 :} write a table of particle data, that you then can edit at leisure. For ordinary listing of decay data, {\tt LULIST(12)} should be used, but that listing could not be read back in by the program. \\ For each compressed flavour code KC = 1--500, one line is written containing KC ({\tt I5}), the basic particle name (i.e. excluding charge etc.) ({\tt 2X,A8}) in {\tt CHAF}, the electric ({\tt I3}), colour charge ({\tt I3}) and particle/antiparticle distinction ({\tt I3}) codes in {\tt KCHG}, the mass ({\tt F12.5}), the mass width ({\tt F12.5}), maximum broadening ({\tt F12.5}) and average invariant lifetime ({\tt 2X,F12.5}) in {\tt PMAS}, and the on/off decay switch (I3) in {\tt MDCY(KC,1).} \\ After a KC line follows one line for each possible decay channel, containing the {\tt MDME} codes ({\tt 5X,2I5}), the branching ratio ({\tt 5X,F12.5}) in {\tt BRAT}, and the {\tt KFDP} codes for the decay products ({\tt 5I8}), with trailing 0's if the number of decay products is smaller than 5. \iteme{= 2 :} read in particle data, as written with {\tt =1} and thereafter edited by you, and use this data subsequently in the current run. Reading is done with fixed format, which means that you have to preserve the format codes described for {\tt =1} during the editing. A number of checks will be made to see if input looks reasonable, with warnings if not. If some decay channel is said not to conserve charge, it should be taken seriously. Warnings that decay is kinematically unallowed need not be as serious, since that particular decay mode may not be switched on unless the particle mass is increased. \iteme{= 3 :} write current particle data as data lines, which can be edited into {\tt BLOCK DATA LUDATA} for a permanent replacement of the particle data. This option is intended for the program author only, not for you. \end{subentry} \iteme{LFN :} the file number which the data should be written to or read from. You must see to it that this file is properly opened for read or write (since the definition of file names is machine dependent). \end{entry} \drawbox{KK = KLU(I,J)}\label{p:KLU} \begin{entry} \itemc{Purpose:} to provide various integer-valued event data. Note that many of the options available (in particular {\tt I}$ > 0$, {\tt J}$\geq 14$) which refer to event history will not work after a {\tt LUEDIT} call. \iteme{I=0, J= :} properties referring to the complete event. \begin{subentry} \iteme{= 1 :} {\tt N}, total number of lines in event record. \iteme{= 2 :} total number of partons/particles remaining after fragmentation and decay. \iteme{= 6 :} three times the total charge of remaining (stable) partons and particles. \end{subentry} \iteme{I>0, J= :} properties referring to the entry in line no. {\tt I} of the event record. \begin{subentry} \iteme{= 1 - 5 :} {\tt K(I,1) - K(I,5)}, i.e. parton/particle status KS, flavour code KF and origin/decay product/colour-flow information. \iteme{= 6 :} three times parton/particle charge. \iteme{= 7 :} 1 for a remaining entry, 0 for a decayed, fragmented or documentation entry. \iteme{= 8 :} KF code ({\tt K(I,2)}) for a remaining entry, 0 for a decayed, fragmented or documentation entry. \iteme{= 9 :} KF code ({\tt K(I,2)}) for a parton (i.e. not colour neutral entry), 0 for a particle. \iteme{= 10 :} KF code ({\tt K(I,2)}) for a particle (i.e. colour neutral entry), 0 for a parton. \iteme{= 11 :} compressed flavour code KC. \iteme{= 12 :} colour information code, i.e. 0 for colour neutral, 1 for colour triplet, -1 for antitriplet and 2 for octet. \iteme{= 13 :} flavour of `heaviest' quark or antiquark (i.e. with largest code) in hadron or diquark (including sign for antiquark), 0 else. \iteme{= 14 :} generation number. Beam particles or virtual exchange particles are generation 0, original jets/particles generation 1 and then 1 is added for each step in the fragmentation/decay chain. \iteme{= 15 :} line number of ancestor, i.e. predecessor in first generation (generation 0 entries are disregarded). \iteme{= 16 :} rank of a hadron in the jet it belongs to. Rank denotes the ordering in flavour space, with hadrons containing the original flavour of the jet having rank 1, increasing by 1 for each step away in flavour ordering. All decay products inherit the rank of their parent. Whereas the meaning of a first-rank hadron in a quark jet is always well-defined, the definition of higher ranks is only meaningful for independently fragmenting quark jets. In other cases, rank refers to the ordering in the actual simulation, which may be of little interest. \iteme{= 17 :} generation number after a collapse of a jet system into one particle, with 0 for an entry not coming from a collapse, and -1 for entry with unknown history. A particle formed in a collapse is generation 1, and then one is added in each decay step. \iteme{= 18 :} number of decay/fragmentation products (only defined in a collective sense for fragmentation). \iteme{= 19 :} origin of colour for showering parton, 0 else. \iteme{= 20 :} origin of anticolour for showering parton, 0 else. \iteme{= 21 :} position of colour daughter for showering parton, 0 else. \iteme{= 22 :} position of anticolour daughter for showering parton, 0 else. \end{subentry} \end{entry} \drawbox{PP = PLU(I,J)}\label{p:PLU} \begin{entry} \itemc{Purpose:} to provide various real-valued event data. Note that some of the options available ({\tt I}$ > 0$, {\tt J}$ = 20$--25), which are primarily intended for studies of systems in their respective c.m. frame, requires that a {\tt LUEXEC} call has been made for the current initial parton/particle configuration, but that the latest {\tt LUEXEC} call has not been followed by a {\tt LUROBO} one. \iteme{I=0, J= :} properties referring to the complete event. \begin{subentry} \iteme{= 1 - 4 :} sum of $p_x$, $p_y$, $p_z$ and $E$, respectively, for the stable remaining entries. \iteme{= 5 :} invariant mass of the stable remaining entries. \iteme{= 6 :} sum of electric charge of the stable remaining entries. \end{subentry} \iteme{I>0, J= :} properties referring to the entry in line no. {\tt I} of the event record. \begin{subentry} \iteme{= 1 - 5 :} {\tt P(I,1) - P(I,5)}, i.e. normally $p_x$, $p_y$, $p_z$, $E$ and $m$ for jet/particle. \iteme{= 6 :} electric charge $e$. \iteme{= 7 :} squared momentum $|\mbf{p}|^2 = p_x^2 + p_y^2 + p_z^2$. \iteme{= 8 :} absolute momentum $|\mbf{p}|$. \iteme{= 9 :} squared transverse momentum $\pT^2 = p_x^2 + p_y^2$. \iteme{= 10 :} transverse momentum $\pT$. \iteme{= 11 :} squared transverse mass $m_{\perp}^2 = m^2 + p_x^2 + p_y^2$. \iteme{= 12 :} transverse mass $m_{\perp}$. \iteme{= 13 - 14 :} polar angle $\theta$ in radians (between 0 and $\pi$) or degrees, respectively. \iteme{= 15 - 16 :} azimuthal angle $\varphi$ in radians (between $-\pi$ and $\pi$) or degrees, respectively. \iteme{= 17 :} true rapidity $y = (1/2) \, \ln((E+p_z)/(E-p_z))$. \iteme{= 18 :} rapidity $y_{\pi}$ obtained by assuming that the particle is a pion when calculating the energy $E$, to be used in the formula above, from the (assumed known) momentum $\mbf{p}$. \iteme{= 19 :} pseudorapidity $\eta = (1/2) \, \ln((p+p_z)/(p-p_z))$. \iteme{= 20 :} momentum fraction $x_p = 2|\mbf{p}|/W$, where $W$ is the total energy of initial jet/particle configuration. \iteme{= 21 :} $x_F = 2p_z/W$ (Feynman-$x$ if system is studied in c.m. frame). \iteme{= 22 :} $x_{\perp} = 2\pT/W$. \iteme{= 23 :} $x_E = 2E/W$. \iteme{= 24 :} $z_+ = (E+p_z)/W$. \iteme{= 25 :} $z_- = (E-p_z)/W$. \end{subentry} \end{entry} \subsection{The General Switches and Parameters} \label{ss:JETswitch} The common block {\tt LUDAT1} is, next to {\tt LUJETS}, the one a {\Je} user is most likely to access. Here one may control in detail what the program is to do, if the default mode of operation is not satisfactory. \drawbox{COMMON/LUDAT1/MSTU(200),PARU(200),MSTJ(200),PARJ(200)}% \label{p:LUDAT1} \begin{entry} \itemc{Purpose:} to give access to a number of status codes and parameters which regulate the performance of the program as a whole. Here {\tt MSTU} and {\tt PARU} are related to utility functions, as well as a few parameters of the Standard Model, while {\tt MSTJ} and {\tt PARJ} affect the underlying physics assumptions. Some of the variables in {\tt LUDAT1} are described elsewhere, and are therefore here only reproduced as references to the relevant sections. This in particular applies to many coupling constants mainly used by {\Py}, which are found just after this, in section \ref{ss:coupcons}. \boxsep \iteme{MSTU(1),MSTU(2) :}\label{p:MSTU} (D=0,0) can be used to replace the ordinary lower and upper limits (normally 1 and {\tt N}) for the action of {\tt LUROBO}, and most {\tt LUEDIT} and {\tt LULIST} calls. Are reset to 0 in a {\tt LUEXEC} call. \iteme{MSTU(3) :} (D=0) number of lines with extra information added after line {\tt N}. Is reset to 0 in a {\tt LUEXEC} call, or in an {\tt LUEDIT} call when particles are removed. \iteme{MSTU(4) :} (D=4000) number of lines available in the common block {\tt LUJETS}. Should always be changed if the dimensions of the {\tt K} and {\tt P} arrays are changed by the user, but should otherwise never be touched. Maximum allowed value is 10000, unless {\tt MSTU(5)} is also changed. \iteme{MSTU(5) :} (D=10000) is used in building up the special colour-flow information stored in {\tt K(I,4)} and {\tt K(I,5)} for {\tt K(I,3)=} 3, 13 or 14. The generic form for {\tt j=} 4 or 5 is \\ {\tt K(I,j)}$ = 2 \times${\tt MSTU(5)}$^2 \times$MCFR$ + ${\tt MSTU(5)}$^2 \times$MCTO$ + ${\tt MSTU(5)}$\times$ICFR$ + $ICTO, \\ with notation as in section \ref{ss:evrec}. One should always have {\tt MSTU(5)}$\geq${\tt MSTU(4)}. On a 32 bit machine, values {\tt MSTU(5)}$> 20000$ may lead to overflow problems, and should be avoided. \iteme{MSTU(6) :} (D=500) number of KC codes available in the {\tt KCHG}, {\tt PMAS}, {\tt MDCY}, and {\tt CHAF} arrays; should be changed if these dimensions are changed. \iteme{MSTU(7) :} (D=2000) number of decay channels available in the {\tt MDME}, {\tt BRAT} and {\tt KFDP} arrays; should be changed if these dimensions are changed. \iteme{MSTU(10) :} (D=2) use of parton/particle masses in filling routines ({\tt LU1ENT}, {\tt LU2ENT}, {\tt LU3ENT}, {\tt LU4ENT}). \begin{subentry} \iteme{= 0 :} assume the mass to be zero. \iteme{= 1 :} keep the mass value stored in {\tt P(I,5)}, whatever it is. (This may be used e.g. to describe kinematics with off-mass-shell partons). \iteme{= 2 :} find masses according to mass tables as usual. \end{subentry} \iteme{MSTU(11) :} (D=6) file number to which all program output is directed. It is your responsibility to see to it that the corresponding file is also opened for output. \iteme{MSTU(12) :} (D=1) writing of title page (version number and last date of change for {\Py} and {\Je}) on output file. \begin{subentry} \iteme{= 0 :} not done. \iteme{= 1 :} title page is written at first occasion, at which time {\tt MSTU(12)} is set =0. \end{subentry} \iteme{MSTU(13) :} (D=1) writing of information on variable values changed by a {\tt LUGIVE} call. \begin{subentry} \iteme{= 0 :} no information is provided. \iteme{= 1 :} information is written to standard output. \end{subentry} \iteme{MSTU(14) :} (D=0) if non-zero, this gives the maximum flavour for which a {\tt LULIST(12)} call will give particle data on possible hadrons. With {\tt MSTU(14)=5} only known hadrons, i.e. up to bottom, are listed. If {\tt =0}, only separately specified particles are listed (i.e. either KF$\leq 100$ or else both KF$> 100$ and KC$> 100$). \iteme{MSTU(15) :} (D=1) selection for characters used in particle names to denote an antiparticle; appear in {\tt LULIST} listings or other {\tt LUNAME} applications. \begin{subentry} \iteme{= 1 :} the tilde character `$\sim$'. \iteme{= 2 :} the characters `bar'. \end{subentry} \iteme{MSTU(16) :} (D=1) choice of mother pointers for the particles produced by a fragmenting parton system. \begin{subentry} \iteme{= 1 :} all primary particles of a system point to a line with KF = 92 or 93, for string or independent fragmentation, respectively, or to a line with KF = 91 if a jet system has so small a mass that it is forced to decay into one or two particles. The two (or more) shower initiators of a showering parton system point to a line with KF = 94. The entries with KF = 91--94 in their turn point back to the predecessor partons, so that the KF = 91--94 entries form a part of the event history proper. \iteme{= 2 :} although the lines with KF = 91--94 are present, and contain the correct mother and daughter pointers, they are not part of the event history proper, in that particles produced in string fragmentation point directly to either of the two endpoint partons of the string (depending on the side they were generated from), particles produced in independent fragmentation point to the respective parton they were generated from, particles in small mass systems point to either endpoint parton, and shower initiators point to the original on-mass-shell counterparts. Also the daugher pointers bypass the KF = 91--94 entries. In independent fragmentation, a parton need not produce any particles at all, and then have daughter pointers 0. \itemc{Note :} {\tt MSTU(16)} should not be changed between the generation of an event and the translation of this event record with a {\tt LUHEPC} call, since this may give an erroneous translation of the event history. \end{subentry} \iteme{MSTU(17) :} (D=0) storage option for {\tt MSTU(90)} and associated information on $z$ values for heavy-flavour production. \begin{subentry} \iteme{= 0 :} {\tt MSTU(90)} is reset to zero at each {\tt LUEXEC} call. This is the appropriate course if {\tt LUEXEC} is only called once per event, as is normally the case when you do not yourself call {\tt LUEXEC}. \iteme{= 1 :} you have to reset {\tt MSTU(90)} to zero yourself before each new event. This is the appropriate course if several {\tt LUEXEC} calls may appear for one event, i.e. if you call {\tt LUEXEC} directly. \end{subentry} \iteme{MSTU(19) :} (D=0) advisory warning for unphysical flavour setups in {\tt LU2ENT}, {\tt LU3ENT} or {\tt LU4ENT} calls. \begin{subentry} \iteme{= 0 :} yes. \iteme{= 1 :} no; {\tt MSTU(19)} is reset to 0 in such a call. \end{subentry} \iteme{MSTU(21) :} (D=2) check on possible errors during program execution. Obviously no guarantee is given that all errors will be caught, but some of the most trivial user-caused errors may be found. \begin{subentry} \iteme{= 0 :} errors do not cause any immediate action, rather the program will try to cope, which may mean e.g. that it runs into an infinite loop. \iteme{= 1 :} parton/particle configurations are checked for possible errors. In case of problem, an exit is made from the misbehaving subprogram, but the generation of the event is continued from there on. For the first {\tt MSTU(22)} errors a a message is printed; after that no messages appear. \iteme{= 2 :} parton/particle configurations are checked for possible errors. In case of problem, an exit is made from the misbehaving subprogram, and subsequently from {\tt LUEXEC}. You may then choose to correct the error, and continue the execution by another {\tt LUEXEC} call. For the first {\tt MSTU(22)} errors a message is printed, after that the last event is printed and execution is stopped. \end{subentry} \iteme{MSTU(22) :} (D=10) maximum number of errors that are printed. \iteme{MSTU(23) :} (I) count of number of errors experienced to date. \iteme{MSTU(24) :} (R) type of latest error experienced; reason that event was not generated in full. Is reset at each {\tt LUEXEC} call. \begin{subentry} \iteme{= 0 :} no error experienced. \iteme{= 1 :} have reached end of or are writing outside {\tt LUJETS} memory. \iteme{= 2 :} unknown flavour code or unphysical combination of codes; may also be caused by erroneous string connection information. \iteme{= 3 :} energy or mass too small or unphysical kinematical variable setup. \iteme{= 4 :} program is caught in an infinite loop. \iteme{= 5 :} momentum, energy or charge was not conserved (even allowing for machine precision errors, see {\tt PARU(11)}); is evaluated only after event has been generated in full, and does not apply when independent fragmentation without momentum conservation was used. \iteme{= 6 :} error call from outside the fragmentation/decay package (e.g. the $\ee$ routines). \iteme{= 7 :} inconsistent particle data input in {\tt LUUPDA} ({\tt MUPDA = 2}) or other {\tt LUUPDA}-related problem. \iteme{= 8 :} problems in more peripheral service routines. \iteme{= 9 :} various other problems. \end{subentry} \iteme{MSTU(25) :} (D=1) printing of warning messages. \begin{subentry} \iteme{= 0 :} no warnings are written. \iteme{= 1 :} first {\tt MSTU(26)} warnings are printed, thereafter no warnings appear. \end{subentry} \iteme{MSTU(26) :} (D=10) maximum number of warnings that are printed. \iteme{MSTU(27) :} (I) count of number of warnings experienced to date. \iteme{MSTU(28) :} (R) type of latest warning given, with codes paralleling those for {\tt MSTU(24)}, but of a less serious nature. \iteme{MSTU(31) :} (I) number of {\tt LUEXEC} calls in present run. \iteme{MSTU(32) :} (I) number of entries stored with {\tt LUEDIT(-1)} call. \iteme{MSTU(33) :} (I) if set 1 before a {\tt LUDBRB} call, the {\tt V} vectors (in the particle range to be rotated/boosted) are set 0 before the rotation/boost. {\tt MSTU(33)} is set back to 0 in the {\tt LUDBRB} call. Is inactive in a {\tt LUROBO} call. \iteme{MSTU(41) - MSTU(63) :} switches for event-analysis routines, see section \ref{ss:evanrout}. \iteme{MSTU(70) :} (D=0) the number of lines consisting only of equal signs ({\tt ======}) that are inserted in the event listing obtained with {\tt LULIST(1)}, {\tt LULIST(2)} or {\tt LULIST(3)}, so as to distinguish different sections of the event record on output. At most 10 such lines can be inserted; see {\tt MSTU(71) - MSTU(80)}. Is reset at {\tt LUEDIT} calls with arguments 0--5. \iteme{MSTU(71) - MSTU(80) :} line numbers below which lines consisting only of equal signs ({\tt ======}) are inserted in event listings. Only the first {\tt MSTU(70)} of the 10 allowed positions are enabled. \iteme{MSTU(90) :} number of heavy-flavour hadrons (i.e. hadrons containing charm or heavier flavours) produced in current event, for which the positions in the event record are stored in {\tt MSTU(91) - MSTU(98)} and the $z$ values in the fragmentation in {\tt PARU(91) - PARU(98)}. At most eight values will be stored (normally this is no problem). No $z$ values can be stored for those heavy hadrons produced when a string has so small mass that it collapses to one or two particles, nor for those produced as one of the final two particles in the fragmentation of a string. If {\tt MSTU(17)=1}, {\tt MSTU(90)} should be reset to zero by you before each new event, else this is done automatically. \iteme{MSTU(91) - MSTU(98) :} the first {\tt MSTU(90)} positions will be filled with the line numbers of the heavy-flavour hadrons produced in the current event. See {\tt MSTU(90)} for additional comments. Note that the information is corrupted by calls to {\tt LUEDIT} with options 0--5 and 21--23; calls with options 11--15 work, however. \iteme{MSTU(101) - MSTU(118) :} switches related to couplings, see section \ref{ss:coupcons}. \iteme{MSTU(161), MSTU(162) :} information used by event-analysis routines, see section \ref{ss:evanrout}. \iteme{MSTU(181) :} (R) {\Je} version number. \iteme{MSTU(182) :} (R) {\Je} subversion number. \iteme{MSTU(183) :} (R) last year of change for {\Je}. \iteme{MSTU(184) :} (R) last month of change for {\Je}. \iteme{MSTU(185) :} (R) last day of change for {\Je}. \iteme{MSTU(186) :} (R) earliest subversion of {\Py} version 5 with which this {\Je} subversion can be run. \boxsep \iteme{PARU(1) :}\label{p:PARU} (R) $\pi \approx 3.1415927$. \iteme{PARU(2) :} (R) $2\pi \approx 6.2831854$. \iteme{PARU(3) :} (D=0.1973) conversion factor for GeV$^{-1} \to$ fm or fm$^{-1} \to$ GeV. \iteme{PARU(4) :} (D=5.068) conversion factor for fm $\to$ GeV$^{-1}$ or GeV $\to$ fm$^{-1}$. \iteme{PARU(5) :} (D=0.3894) conversion factor for GeV$^{-2} \to$ mb or mb$^{-1} \to$ GeV$^2$. \iteme{PARU(6) :} (D=2.568) conversion factor for mb $\to$ GeV$^{-2}$ or GeV$^2 \to$ mb$^{-1}$. \iteme{PARU(11) :} (D=0.001) relative error, i.e. nonconservation of momentum and energy divided by total energy, that may be attributable to machine precision problems before a physics error is suspected (see {\tt MSTU(24)=5}). \iteme{PARU(12) :} (D=0.09 GeV$^2$) effective cut-off in squared mass, below which partons may be recombined to simplify (machine precision limited) kinematics of string fragmentation. \iteme{PARU(13) :} (D=0.01) effective angular cut-off in radians for recombination of partons, used in conjunction with {\tt PARU(12)}. \iteme{PARU(21) :} (I) contains the total energy $W$ of all first generation jets/particles after a {\tt LUEXEC} call; to be used by the {\tt PLU} function for {\tt I>0}, {\tt J=} 20--25. \iteme{PARU(41) - PARU(63) :} parameters for event-analysis routines, see section \ref{ss:evanrout}. \iteme{PARU(91) - PARU(98) :} the first {\tt MSTU(90)} positions will be filled with the fragmentation $z$ values used internally in the generation of heavy-flavour hadrons --- how these are translated into the actual energies and momenta of the observed hadrons is a complicated function of the string configuration. The particle with $z$ value stored in {\tt PARU(i)} is to be found in line {\tt MSTU(i)} of the event record. See {\tt MSTU(90)} and {\tt MSTU(91) - MSTU(98)} for additional comments. \iteme{PARU(101) - PARU(195) :} various coupling constants and parameters related to couplings, see section \ref{ss:coupcons}. \boxsep \iteme{MSTJ(1) :}\label{p:MSTJ} (D=1) choice of fragmentation scheme. \begin{subentry} \iteme{= 0 :} no jet fragmentation at all. \iteme{= 1 :} string fragmentation according to the Lund model. \iteme{= 2 :} independent fragmentation, according to specification in {\tt MSTJ(2)} and {\tt MSTJ(3)}. \end{subentry} \iteme{MSTJ(2) :} (D=3) gluon jet fragmentation scheme in independent fragmentation. \begin{subentry} \iteme{= 1 :} a gluon is assumed to fragment like a random $\d$, $\u$ or $\s$ quark or antiquark. \iteme{= 2 :} as {\tt =1}, but longitudinal (see {\tt PARJ(43)}, {\tt PARJ(44)} and {\tt PARJ(59)}) and transverse (see {\tt PARJ(22)}) momentum properties of quark or antiquark substituting for gluon may be separately specified. \iteme{= 3 :} a gluon is assumed to fragment like a pair of a $\d$, $\u$ or $\s$ quark and its antiquark, sharing the gluon energy according to the Altarelli-Parisi splitting function. \iteme{= 4 :} as {\tt =3}, but longitudinal (see {\tt PARJ(43)}, {\tt PARJ(44)} and {\tt PARJ(59)}) and transverse (see {\tt PARJ(22)}) momentum properties of quark and antiquark substituting for gluon may be separately specified. \end{subentry} \iteme{MSTJ(3) :} (D=0) energy, momentum and flavour conservation options in independent fragmentation. Whenever momentum conservation is described below, energy and flavour conservation is also implicitly assumed. \begin{subentry} \iteme{= 0 :} no explicit conservation of any kind. \iteme{= 1 :} particles share momentum imbalance compensation according to their energy (roughly equivalent to boosting event to c.m. frame). This is similar to the approach in the Ali et al. program \cite{Ali80}. \iteme{= 2 :} particles share momentum imbalance compensation according to their longitudinal mass with respect to the imbalance direction. \iteme{= 3 :} particles share momentum imbalance compensation equally. \iteme{= 4 :} transverse momenta are compensated separately within each jet, longitudinal momenta are rescaled so that ratio of final jet to initial parton momentum is the same for all the jets of the event. This is similar to the approach in the Hoyer et al. program \cite{Hoy79}. \iteme{= 5 :} only flavour is explicitly conserved. \iteme{= 6 - 10 :} as {\tt =1 - 5}, except that above several colour singlet systems that followed immediately after each other in the event listing (e.g. $\q\qbar\q\qbar$) were treated as one single system, whereas here they are treated as separate systems. \iteme{= -1 :} independent fragmentation, where also particles moving backwards with respect to the jet direction are kept, and thus the amount of energy and momentum mismatch may be large. \end{subentry} \iteme{MSTJ(11) :} (D=4) choice of longitudinal fragmentation function, i.e. how large a fraction of the energy available a newly-created hadron takes. \begin{subentry} \iteme{= 1 :} the Lund symmetric fragmentation function, see {\tt PARJ(41) - PARJ(45)}. \iteme{= 2 :} choice of some different forms for each flavour separately, see {\tt PARJ(51) - PARJ(59)}. \iteme{= 3 :} hybrid scheme, where light flavours are treated with symmetric Lund ({\tt =1}), but charm and heavier can be separately chosen, e.g. according to the SLAC function ({\tt =2}). \iteme{= 4 :} the Lund symmetric fragmentation function ({\tt =1}), for heavy endpoint quarks modified according to the Bowler (Artru--Mennessier, Morris) space--time picture of string evolution, see {\tt PARJ(46)}. \iteme{= 5 :} as {\tt =4}, but with possibility to interpolate between Bowler and Lund separately for $\c$, $\b$ and $\t$; see {\tt PARJ(46) - PARJ(48)}. \end{subentry} \iteme{MSTJ(12) :} (D=2) choice of baryon production model. \begin{subentry} \iteme{= 0 :} no baryon-antibaryon pair production at all; initial diquark treated as a unit. \iteme{= 1 :} diquark-antidiquark pair production allowed; diquark treated as a unit. \iteme{= 2 :} diquark-antidiquark pair production allowed, with possibility for diquark to be split according to the `popcorn' scheme. \iteme{= 3 :} as {\tt =2}, but additionally the production of first rank baryons may be suppressed by a factor {\tt PARJ(19)}. \end{subentry} \iteme{MSTJ(13) :} (D=0) generation of transverse momentum for endpoint quark(s) of single quark jet or $\q\qbar$ jet system (in multijet events no endpoint transverse momentum is ever allowed for). \begin{subentry} \iteme{= 0 :} no transverse momentum for endpoint quarks. \iteme{= 1 :} endpoint quarks obtain transverse momenta like ordinary $\q\qbar$ pairs produced in the field (see {\tt PARJ(21)}); for 2-jet systems the endpoints obtain balancing transverse momenta. \end{subentry} \iteme{MSTJ(14) :} (D=1) treatment of a colour-singlet jet system with a low invariant mass. \begin{subentry} \iteme{= 0 :} no precautions are taken, meaning that problems may occur in {\tt LUSTRF} (or {\tt LUINDF}) later on. \iteme{= 1 :} small jet systems are allowed to collapse into two particles or, failing that, one single particle. Normally all small systems are treated this way, starting with the smallest one, but some systems would require more work and are left untreated; they include diquark-antidiquark pairs below the two-particle threshold. \iteme{= -1 :} special option for {\tt LUPREP} calls, where no precautions are taken (as for {\tt =0}), but, in addition, no checks are made on the presence of small-mass systems; i.e. {\tt LUPREP} only rearranges colour strings. \end{subentry} \iteme{MSTJ(15) :} (D=0) production probability for new flavours. \begin{subentry} \iteme{= 0 :} according to standard Lund parametrization, as given by {\tt PARJ(1) - PARJ(20)}. \iteme{= 1 :} according to probabilities stored in {\tt PARF(201) - PARF(1960)}; note that no default values exist here, i.e. {\tt PARF} must be set by you. The {\tt MSTJ(12)} switch can still be used to set baryon production mode, with the modification that {\tt MSTJ(12)=2} here allows an arbitrary number of mesons to be produced between a baryon and an antibaryon (since the probability for diquark $\to$ meson $+$ new diquark is assumed independent of prehistory). \end{subentry} \iteme{MSTJ(21) :} (D=2) form of particle decays. \begin{subentry} \iteme{= 0 :} all particle decays are inhibited. \iteme{= 1 :} a particle declared unstable in the {\tt MDCY} vector, and with decay channels defined, may decay within the region given by {\tt MSTJ(22)}. A particle may decay into jets, which then fragment further according to the {\tt MSTJ(1)} value. \iteme{= 2 :} as {\tt =1}, except that a $\q\qbar$ jet system produced in a decay (e.g. of a $\B$ meson) is always allowed to fragment according to string fragmentation, rather than according to the {\tt MSTJ(1)} value (this means that momentum, energy and charge are conserved in the decay). \end{subentry} \iteme{MSTJ(22) :} (D=1) cut-off on decay length for a particle that is allowed to decay according to {\tt MSTJ(21)} and the {\tt MDCY} value. \begin{subentry} \iteme{= 1 :} a particle declared unstable is also forced to decay. \iteme{= 2 :} a particle is decayed only if its average invariant lifetime is larger than {\tt PARJ(71)}. \iteme{= 3 :} a particle is decayed only if the decay vertex is within a distance {\tt PARJ(72)} of the origin. \iteme{= 4 :} a particle is decayed only if the decay vertex is within a cylindrical volume with radius {\tt PARJ(73)} in the $xy$-plane and extent to $\pm${\tt PARJ(74)} in the $z$ direction. \end{subentry} \iteme{MSTJ(23) :} (D=1) possibility of having a shower evolving from a $\q\qbar$ pair created as decay products. \begin{subentry} \iteme{= 0 :} never. \iteme{= 1 :} whenever the decay channel matrix-element code is {\tt MDME(IDC,2)=} 22, 23 or 33, the two first decay products (if they are partons) are allowed to shower, like a colour-singlet subsystem, with maximum virtuality given by the invariant mass of the pair. \end{subentry} \iteme{MSTJ(24) :} (D=2) particle masses. \begin{subentry} \iteme{= 0 :} discrete mass values are used. \iteme{= 1 :} particles registered as having a mass width in the {\tt PMAS} vector are given a mass according to a truncated Breit--Wigner shape, linear in $m$, eq. (\ref{dec:BWlin}). \iteme{= 2 :} as {\tt =1}, but gauge bosons (actually all particles with $|$KF$| \leq 100$) are distributed according to a Breit--Wigner quadratic in $m$, as obtained from propagators. \iteme{= 3 :} as {\tt =1}, but Breit--Wigner shape is always quadratic in $m$, eq. (\ref{dec:BWtwo}). \end{subentry} \iteme{MSTJ(25) :} (D=1) inclusion of the $\W^{\pm}$ propagator, in addition to the standard, `infinitely heavy' weak $V-A$ matrix element, in the decay of a $\t$, $\lrm$ or $\hrm$ quark, or $\chi$ lepton. \begin{subentry} \iteme{= 0 :} not included. \iteme{= 1 :} included. \end{subentry} \iteme{MSTJ(26) :} (D=2) inclusion of $\B$--$\br{\B}$ mixing in decays. \begin{subentry} \iteme{= 0 :} no. \iteme{= 1 :} yes, with mixing parameters given by {\tt PARJ(76)} and {\tt PARJ(77)}. Mixing decays are not specially marked. \iteme{= 2 :} yes, as {\tt =1}, but a $\B$ ($\br{\B}$) that decays as a $\br{\B}$ ($\B$) is marked as {\tt K(I,1)=12} rather than the normal {\tt K(I,1)=11}. \end{subentry} \iteme{MSTJ(27) :} (D=2) possibility for the $\b$ quark to develop a shower in the decay of a top hadron, i.e. $\T \to \W^+ \b \qbar$, where $\qbar$ is a spectator quark. \begin{subentry} \iteme{= 0 :} no, i.e. $\b$ jet is narrow, low-multiplicity. \iteme{= 1 :} the $\b$ is allowed to shower and the $\W$ momentum (in the rest frame of the $\T$) is reduced acccordingly. The $\W$ is therafter assumed to decay isotropically. \iteme{= 2 :} the $\b$ is allowed to shower, as in {\tt =1}, but the $\W$ decays anisotropically, with the same polarization as in the standard weak decay of option {\tt =0}. In principle this is better than option {\tt =1}, but in practice there is no big difference. \end{subentry} \iteme{MSTJ(41) - MSTJ(50) :} switches for time-like parton showers, see section \ref{ss:showrout}. \iteme{MSTJ(51) :} (D=0) inclusion of Bose--Einstein effects. \begin{subentry} \iteme{= 0 :} no effects included. \iteme{= 1 :} effects included according to an exponential parametrization $C_2(Q) = 1 + ${\tt PARJ(92)}$\times \exp(-Q/${\tt PARJ(93)}$)$, where $C_2(Q)$ represents the ratio of particle production at $Q$ with Bose--Einstein effects to that without, and the relative momentum $Q$ is defined by $Q^2(p_1,p_2) = -(p_1 - p_2)^2 = (p_1 + p_2)^2 - 4m^2$. Particles with width broader than {\tt PARJ(91)} are assumed to have time to decay before Bose--Einstein effects are to be considered. \iteme{= 2 :} effects included according to a Gaussian parametrization $C_2(Q) = 1 + ${\tt PARJ(92)}$\times \exp(- (Q/${\tt PARJ(93)}$)^2 )$, with notation and comments as above. \end{subentry} \iteme{MSTJ(52) :} (D=3) number of particle species for which Bose--Einstein correlations are to be included, ranged along the chain $\pi^+$, $\pi^-$, $\pi^0$, $\K^+$, $\K^-$, $\K^0_{\mrm{S}}$, $\K^0_{\mrm{L}}$, $\eta$ and $\eta'$. Default corresponds to including all pions ($\pi^+$, $\pi^-$, $\pi^0$), 7 to including all Kaons as well, and 9 is maximum. \iteme{MSTJ(91) :} (I) flag when generating gluon jet with options {\tt MSTJ(2)=} 2 or 4 (then {\tt =1}, else {\tt =0}). \iteme{MSTJ(92) :} (I) flag that a $\q\qbar$ or $\g\g$ pair or a $\g\g\g$ triplet created in {\tt LUDECY} should be allowed to shower, is 0 if no pair or triplet, is the entry number of the first parton if a pair indeed exists, is the entry number of the first parton, with a $-$ sign, if a triplet indeed exists. \iteme{MSTJ(93) :} (I) switch for {\tt ULMASS} action. Is reset to 0 in {\tt ULMASS} call. \begin{subentry} \iteme{= 0 :} ordinary action. \iteme{= 1 :} light ($\d$, $\u$, $\s$, $\c$, $\b$) quark masses are taken from {\tt PARF(101) - PARF(105)} rather than {\tt PMAS(1,1) - PMAS(5,1)}. Diquark masses are given as sum of quark masses, without spin splitting term. \iteme{= 2 :} as {\tt =1}. Additionally the constant terms {\tt PARF(121)} and {\tt PARF(122)} are subtracted from quark and diquark masses, respectively. \end{subentry} \iteme{MSTJ(101) - MSTJ(121) :} switches for $\ee$ event generation, see section \ref{ss:eeroutines}. \boxsep \iteme{PARJ(1) :}\label{p:PARJ} (D=0.10) is ${\cal P}(\q\q)/{\cal P}(\q)$, the suppression of diquark-antidiquark pair production in the colour field, compared with quark--antiquark production. \iteme{PARJ(2) :} (D=0.30) is ${\cal P}(\s)/{\cal P}(u)$, the suppression of $\s$ quark pair production in the field compared with $\u$ or $\d$ pair production. \iteme{PARJ(3) :} (D=0.4) is $({\cal P}(\u\s)/{\cal P}(\u\d))/({\cal P}(\s)/{\cal P}(\d))$, the extra suppression of strange diquark production compared with the normal suppression of strange quarks. \iteme{PARJ(4) :} (D=0.05) is $(1/3){\cal P}(\u\d_1)/{\cal P}(\u\d_0)$, the suppression of spin 1 diquarks compared with spin 0 ones (excluding the factor 3 coming from spin counting). \iteme{PARJ(5) :} (D=0.5) parameter determining relative occurence of baryon production by $BM\br{B}$ and by $B\br{B}$ configurations in the popcorn baryon production model, roughly ${\cal P}(BM\br{B})/({\cal P}(B\br{B})+{\cal P}(BM\br{B})) =$ {\tt PARJ(5)}$/(0.5+${\tt PARJ(5)}$)$. \iteme{PARJ(6) :} (D=0.5) extra suppression for having a $\s\sbar$ pair shared by the $B$ and $\br{B}$ of a $BM\br{B}$ situation. \iteme{PARJ(7) :} (D=0.5) extra suppression for having a strange meson $M$ in a $BM\br{B}$ configuration. \iteme{PARJ(11) - PARJ(17) :} parameters that determine the spin of mesons. \begin{subentry} \iteme{PARJ(11) :} (D=0.5) is the probability that a light meson (containing $\u$ and $\d$ quarks only) has spin 1 (with {\tt 1-PARJ(11)} the probability for spin 0) when formed in fragmentation. \iteme{PARJ(12) :} (D=0.6) is the probability that a strange meson has spin 1. \iteme{PARJ(13) :} (D=0.75) is the probability that a charm or heavier meson has spin 1. \iteme{PARJ(14) :} (D=0.) is the probability that a spin = 0 meson is produced with an orbital angular momentum 1, for a total spin = 1. \iteme{PARJ(15) :} (D=0.) is the probability that a spin = 1 meson is produced with an orbital angular momentum 1, for a total spin = 0. \iteme{PARJ(16) :} (D=0.) is the probability that a spin = 1 meson is produced with an orbital angular momentum 1, for a total spin = 1. \iteme{PARJ(17) :} (D=0.) is the probability that a spin = 1 meson is produced with an orbital angular momentum 1, for a total spin = 2. \itemc{Note :} the end result of the numbers above is that, with {\tt i} = 11, 12 or 13, depending on flavour content, \\ ${\cal P}(S = 0, L = 0, J = 0) = (1 - \mtt{PARJ(i)}) \times (1 - \mtt{PARJ(14)})$, \\ ${\cal P}(S = 0, L = 1, J = 1) = (1 - \mtt{PARJ(i)}) \times \mtt{PARJ(14)}$, \\ ${\cal P}(S = 1, L = 0, J = 1) = $ \\ \mbox{~}\hfill$\mtt{PARJ(i)} \times (1 - \mtt{PARJ(15)} - \mtt{PARJ(16)} - \mtt{PARJ(17)})$, \\ ${\cal P}(S = 1, L = 1, J = 0) = \mtt{PARJ(i)} \times \mtt{PARJ(15)}$, \\ ${\cal P}(S = 1, L = 1, J = 1) = \mtt{PARJ(i)} \times \mtt{PARJ(16)}$, \\ ${\cal P}(S = 1, L = 1, J = 2) = \mtt{PARJ(i)} \times \mtt{PARJ(17)}$, \\ where $S$ is the quark `true' spin and $J$ is the total spin, usually called the spin $s$ of the meson. \end{subentry} \iteme{PARJ(18) :} (D=1.) is an extra suppression factor multiplying the ordinary {\bf SU(6)} weight for spin $3$/2 baryons, and hence a means to break {\bf SU(6)} in addition to the dynamic breaking implied by {\tt PARJ(2)}, {\tt PARJ(3)}, {\tt PARJ(4)}, {\tt PARJ(6)} and {\tt PARJ(7)}. \iteme{PARJ(19) :} (D=1.) extra baryon suppression factor, which multiplies the ordinary diquark-antidiquark production probability for the breakup closest to the endpoint of a string, but leaves other breaks unaffected. Is only used for {\tt MSTJ(12)=3}. \iteme{PARJ(21) :} (D=0.36 GeV) corresponds to the width $\sigma$ in the Gaussian $p_x$ and $p_y$ transverse momentum distributions for primary hadrons. See also {\tt PARJ(22) - PARJ(24)}. \iteme{PARJ(22) :} (D=1.) relative increase in transverse momentum in a gluon jet generated with {\tt MSTJ(2)=} 2 or 4. \iteme{PARJ(23), PARJ(24) :} (D=0.01, 2.) a fraction {\tt PARJ(23)} of the Gaussian transverse momentum distribution is taken to be a factor {\tt PARJ(24)} larger than input in {\tt PARJ(21)}. This gives a simple parametrization of non-Gaussian tails to the Gaussian shape assumed above. \iteme{PARJ(25) :} (D=1.) extra suppression factor for $\eta$ production in fragmentation; if an $\eta$ is rejected a new flavour pair is generated and a new hadron formed. \iteme{PARJ(26) :} (D=0.4) extra suppression factor for $\eta'$ production in fragmentation; if an $\eta'$ is rejected a new flavour pair is generated and a new hadron formed. \iteme{PARJ(31) :} (D=0.1 GeV) gives the remaining $W_+$ below which the generation of a single jet is stopped (it is chosen smaller than a pion mass, so that no hadrons moving in the forward direction are missed). \iteme{PARJ(32) :} (D=1. GeV) is, with quark masses added, used to define the minimum allowable energy of a colour-singlet jet system. \iteme{PARJ(33) - PARJ(34) :} (D=0.8 GeV, 1.5 GeV) are, together with quark masses, used to define the remaining energy below which the fragmentation of a jet system is stopped and two final hadrons formed. {\tt PARJ(33)} is normally used, except for {\tt MSTJ(11)=2}, when {\tt PARJ(34)} is used. \iteme{PARJ(36) :} (D=2.) represents the dependence on the mass of the final quark pair for defining the stopping point of the fragmentation. Is strongly correlated to the choice of {\tt PARJ(33) - PARJ(35)}. \iteme{PARJ(37) :} (D=0.2) relative width of the smearing of the stopping point energy. \iteme{PARJ(38) - PARJ(39) :} (D=2.5, 0.6) refers to the probability for reverse rapidity ordering of the final two hadrons, according to eq. (\ref{fr:revord}), where $d_0 = ${\tt PARJ(39)} for {\tt MSTJ(11)}$\neq 2$, and $d = ${\tt PARJ(39)} for {\tt MSTJ(11)=2}. \iteme{PARJ(41), PARJ(42) :} (D=0.3, 0.58 GeV$^{-2}$) give the $a$ and $b$ parameters of the symmetric Lund fragmentation function for {\tt MSTJ(11)=}1, 4 and 5 (and {\tt MSTJ(11)=3} for ordinary hadrons). \iteme{PARJ(43), PARJ(44) :} (D=0.5, 0.9 GeV$^{-2}$) give the $a$ and $b$ parameters as above for the special case of a gluon jet generated with IF and {\tt MSTJ(2)=} 2 or 4. \iteme{PARJ(45) :} (D=0.5) the amount by which the effective $a$ parameter in the Lund flavour dependent symmetric fragmentation function is assumed to be larger than the normal $a$ when diquarks are produced. More specifically, referring to eq. (\ref{fr:LSFFlong}), $a_{\alpha} = ${\tt PARJ(41)} when considering the fragmentation of a quark and = {\tt PARJ(41) + PARJ(45)} for the fragmentation of a diquark, with corresponding expression for $a_{\beta}$ depending on whether the newly created object is a quark or diquark (for an independent gluon jet generated with {\tt MSTJ(2)=} 2 or 4, replace {\tt PARJ(41)} by {\tt PARJ(43)}). In the popcorn model, a meson created in between the baryon and antibaryon has $a_{\alpha} = a_{\beta} = ${\tt PARJ(41) + PARJ(45)}. \iteme{PARJ(46) - PARJ(48) :} (D=3*1.) modification of the Lund symmetric fragmentation for heavy endpoint quarks according to the recipe by Bowler, available when {\tt MSTJ(11)=} 4 or 5 is selected. The shape is given by eq. (\ref{fr:LSFFBowler}). If {\tt MSTJ(11)=4} then $r_{\Q} = ${\tt PARJ(46)} for all flavours, while if {\tt MSTJ(11)=5} then $r_{\c} = ${\tt PARJ(46)}, $r_{\b} = ${\tt PARJ(47)} and $r_{\Q} = ${\tt PARJ(48)} for $\t$ and heavier. {\tt PARJ(46) - PARJ(48)} thus provide a possibility to interpolate between the `pure' Bowler shape, $r = 1$, and the normal Lund one, $r = 0$. The additional modifications made in {\tt PARJ(43) - PARJ(45)} are automatically taken into account, if necessary. \iteme{PARJ(51) - PARJ(58) :} (D=3*0.77, $-0.05$, $-0.005$, 3*$-0.00001$) give a choice of four possible ways to parametrize the fragmentation function for {\tt MSTJ(11)=2} (and {\tt MSTJ(11)=3} for charm and heavier). The fragmentation of each flavour KF may be chosen separately; for a diquark the flavour of the heaviest quark is used. With $c = ${\tt PARJ(50+KF)}, the parametrizations are: \\ $0 \leq c \leq 1$ : Field-Feynman, $f(z) = 1 - c + 3c(1-z)^2$; \\ $-1 \leq c < 0$ : SLAC, $f(z) = 1/(z(1-1/z-(-c)/(1-z))^2)$; \\ $c > 1$ : power peaked at $z=0$, $f(z) = (1-z)^{c-1}$; \\ $c < -1$ : power peaked at $z=1$, $f(z) = z^{-c-1}$. \iteme{PARJ(59) :} (D=1.) replaces {\tt PARJ(51) - PARJ(53)} for gluon jet generated with {\tt MSTJ(2)=} 2 or 4. \iteme{PARJ(61) - PARJ(63) :} (D=4.5, 0.7, 0.) parametrizes the energy dependence of the primary multiplicity distribution in phase-space decays. The former two correspond to $c_1$ and $c_2$ of eq. (\ref{dec:multsel}), while the latter allows a further additive term in the multiplicity specifically for onium decays. \iteme{PARJ(64) :} (0.003 GeV) minimum kinetic energy in decays (safety margin for numerical precision errors). \iteme{PARJ(65) :} (D=0.5 GeV) mass which, in addition to the spectator quark ordiquark mass, is not assumed to partake in the weak decay of a heavy quark in a hadron. \iteme{PARJ(66) :} (D=0.5) relative probability that colour is rearranged when two singlets are to be formed from decay products. Only applies for {\tt MDME(IDC,2)=} 11--30, i.e. low-mass phase-space decays. \iteme{PARJ(71) :} (D=10 mm) maximum average invariant lifetime for particles allowed to decay in the {\tt MSTJ(22)=2} option. With the default value, $\K_{\mrm{S}}^0$, $\Lambda$, $\Sigma^-$, $\Sigma^+$, $\Xi^-$, $\Xi^0$ and $\Omega^-$ are stable (in addition to those normally taken to be stable), but charm and bottom do still decay. \iteme{PARJ(72) :} (D=1000 mm) maximum distance from the origin at which a decay is allowed to take place in the {\tt MSTJ(22)=3} option. \iteme{PARJ(73) :} (D=100 mm) maximum cylindrical distance $\rho = \sqrt{x^2 + y^2}$ from the origin at which a decay is allowed to take place in the {\tt MSTJ(22)=4} option. \iteme{PARJ(74) :} (D=1000 mm) maximum z distance from the origin at which a decay is allowed to take place in the {\tt MSTJ(22)=4} option. \iteme{PARJ(76) :} (D=0.7) mixing parameter $x_d = \Delta M/\Gamma$ in $\B^0$--$\br{\B}^0$ system. \iteme{PARJ(77) :} (D=10.) mixing parameter $x_s = \Delta M/\Gamma$ in $\B_s^0$--$\br{\B}_s^0$ system. \iteme{PARJ(81) - PARJ(88) :} parameters for time-like parton showers, see section \ref{ss:showrout}. \iteme{PARJ(91) :} (D=0.020 GeV) minimum particle width in {\tt PMAS(KC,2)}, above which particle decays are assumed to take place before the stage where Bose--Einstein effects are introduced. \iteme{PARJ(92) :} (D=1.) nominal strength of Bose--Einstein effects for $Q = 0$, see {\tt MSTJ(51)}. This parameter, often denoted $\lambda$, expresses the amount of incoherence in particle production. Due to the simplified picture used for the Bose--Einstein effects, in particular for effects from three nearby identical particles, the actual $\lambda$ of the simulated events may be larger than the input value. \iteme{PARJ(93) :} (D=0.20 GeV) size of the Bose--Einstein effect region in terms of the $Q$ variable, see {\tt MSTJ(51)}. The more conventional measure, in terms of the radius $R$ of the production volume, is given by $R = \hbar/${\tt PARJ(93)}$ \approx 0.2$ fm$\times$GeV/{\tt PARJ(93)}$ = ${\tt PARU(3)}/{\tt PARJ(93)}. \iteme{PARJ(121) - PARJ(171) :} parameters for $\ee$ event generation, see section \ref{ss:eeroutines}. \end{entry} \subsection{Couplings} \label{ss:coupcons} In this section we collect information on the two routines for running $\alphas$ and $\alphaem$, and on other couplings of standard and non-standard particles. Although originally begun for {\Je} applications, this section has rapidly expanded towards the non-standard aspects, and is thus more of interest for {\Py} applications than for {\Je} itself. It could therefore equally well have been put somewhere else in this manual. A few couplings indeed appear in the {\tt PARP} array, see section \ref{ss:PYswitchpar}. \drawbox{ALEM = ULALEM(Q2)}\label{p:ULALEM} \begin{entry} \itemc{Purpose:} to calculate the running electromagnetic coupling constant $\alphaem$. Expressions used are described in ref. \cite{Kle89}. See {\tt MSTU(101)} and {\tt PARU(101)}. \iteme{Q2 :} the momentum transfer scale $Q^2$ at which to evaluate $\alphaem$. \end{entry} \drawbox{ALPS = ULALPS(Q2)}\label{p:ULALPS} \begin{entry} \itemc{Purpose:} to calculate the running strong coupling constant $\alphas$. The first- and second-order expressions are given by eqs. (\ref{ee:aS3j}) and (\ref{ee:aS4j}). See {\tt MSTU(111) - MSTU(118)} and {\tt PARU(111) - PARU(118)} for options. \iteme{Q2 :} the momentum transfer scale $Q^2$ at which to evaluate $\alphas$. \end{entry} \drawbox{COMMON/LUDAT1/MSTU(200),PARU(200),MSTJ(200),PARJ(200)} \begin{entry} \itemc{Purpose:} to give access to a number of status codes and parameters which regulate the performance of the program as a whole. Here only those related to couplings are described; the main description is found in section \ref{ss:JETswitch}. \boxsep \iteme{MSTU(101) :}\label{p:MSTU101} (D=1) procedure for $\alphaem$ evaluation in the {\tt ULALEM} function. \begin{subentry} \iteme{= 0 :} $\alphaem$ is taken fixed at the value {\tt PARU(101)}. \iteme{= 1 :} $\alphaem$ is running with the $Q^2$ scale, taking into account corrections from fermion loops ($\e$, $\mu$, $\tau$, $\d$, $\u$, $\s$, $\c$, $\b$). \end{subentry} \iteme{MSTU(111) :} (D=1) order of $\alphas$ evaluation in the {\tt ULALPS} function. Is overwritten in {\tt LUEEVT}, {\tt LUONIA} or {\tt PYINIT} calls with the value desired for the process under study. \begin{subentry} \iteme{= 0 :} $\alphas$ is fixed at the value {\tt PARU(111)}. \iteme{= 1 :} first-order running $\alphas$ is used. \iteme{= 2 :} second-order running $\alphas$ is used. \end{subentry} \iteme{MSTU(112) :} (D=5) the nominal number of flavours assumed in the $\alphas$ expression, with respect to which $\Lambda$ is defined. \iteme{MSTU(113) :} (D=3) minimum number of flavours that may be assumed in $\alphas$ expression, see {\tt MSTU(112)}. \iteme{MSTU(114) :} (D=5) maximum number of flavours that may be assumed in $\alphas$ expression, see {\tt MSTU(112)}. \iteme{MSTU(115) :} (D=0) treatment of $\alphas$ singularity for $Q^2 \to 0$. \begin{subentry} \iteme{= 0 :} allow it to diverge like $1/\ln(Q^2/\Lambda^2)$. \iteme{= 1 :} soften the divergence to $1/\ln(1 + Q^2/\Lambda^2)$. \iteme{= 2 :} freeze $Q^2$ evolution below {\tt PARU(114)}, i.e. the effective argument is $\max(Q^2, ${\tt PARU(114)}$)$. \end{subentry} \iteme{MSTU(118) :} (I) number of flavours $n_f$ found and used in latest {\tt ULALPS} call. \boxsep \iteme{PARU(101) :}\label{p:PARU101} (D=0.00729735) $\alphaem$, the electromagnetic fine structure constant at vanishing momentum transfer. \iteme{PARU(102) :} (D=0.232) $\ssintw$, the weak mixing angle of the standard electroweak model. \iteme{PARU(108) :} (I) the $\alphaem$ value obtained in the latest call to the {\tt ULALEM} function. \iteme{PARU(111) :} (D=0.20) fix $\alphas$ value assumed in {\tt ULALPS} when {\tt MSTU(111)=0} (and also in parton showers when $\alphas$ is assumed fix there). \iteme{PARU(112) :} (D=0.25 GeV) $\Lambda$ used in running $\alphas$ expression in {\tt ULALPS}. Like {\tt MSTU(111)}, this value is overwritten by the calling physics routines, and is therefore purely nominal. \iteme{PARU(113) :} (D=1.) the flavour thresholds, for the effective number of flavours $n_f$ to use in the $\alphas$ expression, are assumed to sit at $Q^2 = ${\tt PARU(113)}$\times m_{\q}^2$, where $m_{\q}$ is the quark mass. May be overwritten from the calling physics routine. \iteme{PARU(114) :} (D=4 GeV$^2$) $Q^2$ value below which the $\alphas$ value is assumed constant for {\tt MSTU(115)=2}. \iteme{PARU(115) :} (D=10.) maximum $\alphas$ value that {\tt ULALPS} will ever return; is used as a last resort to avoid singularities. \iteme{PARU(117) :} (I) $\Lambda$ value (associated with {\tt MSTU(118)} effective flavours) obtained in latest {\tt ULALPS} call. \iteme{PARU(118) :} (I) $\alphas$ value obtained in latest {\tt ULALPS} call. \iteme{PARU(121) - PARU(130) :} couplings of a new $\Z'^0$; for fermion default values are given by the Standard Model $\Z^0$ values, assuming $\ssintw = 0.23$. Note that e.g. the $\Z'^0$ width contains squared couplings, and thus depends quadratically on the values below. \begin{subentry} \iteme{PARU(121), PARU(122) :} (D=-0.693,-1.) vector and axial couplings of down type quarks to $\Z'^0$. \iteme{PARU(123), PARU(124) :} (D=0.387,1.) vector and axial couplings of up type quarks to $\Z'^0$. \iteme{PARU(125), PARU(126) :} (D=-0.08,-1.) vector and axial couplings of leptons to $\Z'^0$. \iteme{PARU(127), PARU(128) :} (D=1.,1.) vector and axial couplings of neutrinos to $\Z'^0$. \iteme{PARU(129) :} (D=1.) the coupling $Z'^0 \to \W^+ \W^-$ is taken to be {\tt PARU(129)}$\times$(the Standard Model $\Z^0 \to \W^+ \W^-$ coupling)$\times (m_{\W}/m_{\Z'})^2$. This gives a $\Z'^0 \to \W^+ \W^-$ partial width that increases proportionately to the $\Z'^0$ mass. \iteme{PARU(130) :} (D=0.) in the decay chain $\Z'^0 \to \W^+ \W^- \to 4$ fermions, the angular distribution in the $\W$ decays is supposed to be a mixture, with fraction {\tt 1-PARU(130)} corresponding to the same angular distribution between the four final fermions as in $\Z^0 \to \W^+ \W^-$ (mixture of transverse and longitudinal $\W$'s), and fraction {\tt PARU(130)} corresponding to $\H^0 \to \W^+ \W^-$ the same way (longitudinal $\W$'s). \end{subentry} \iteme{PARU(131) - PARU(136) :} couplings of a new $\W'^{\pm}$; for fermions default values are given by the Standard Model $\W^{\pm}$ values (i.e. $V-A$). Note that e.g. the $\W'^{\pm}$ width contains squared couplings, and thus depends quadratically on the values below. \begin{subentry} \iteme{PARU(131), PARU(132) :} (D=1.,-1.) vector and axial couplings of a quark--antiquark pair to $\W'^{\pm}$; is further multiplied by the ordinary CKM factors. \iteme{PARU(133), PARU(134) :} (D=1.,-1.) vector and axial couplings of a lepton-neutrino pair to $\W'^{\pm}$. \iteme{PARU(135) :} (D=1.) the coupling $\W'^{\pm} \to \Z^0 \W^{\pm}$ is taken to be {\tt PARU(135)}$\times$(the Standard Model $\W^{\pm} \to \Z^0 \W^{\pm}$ coupling)$\times (m_{\W}/m_{W'})^2$. This gives a $\W'^{\pm} \to \Z^0 \W^{\pm}$ partial width that increases proportionately to the $\W'$ mass. \iteme{PARU(136) :} (D=0.) in the decay chain $\W'^{\pm} \to \Z^0 \W^{\pm} \to 4$ fermions, the angular distribution in the $\W/\Z$ decays is supposed to be a mixture, with fraction {\tt 1-PARU(130)} corresponding to the same angular distribution between the four final fermions as in $\W^{\pm} \to \Z^0 \W^{\pm}$ (mixture of transverse and longitudinal $\W/\Z$'s), and fraction {\tt PARU(130)} corresponding to $\H^{\pm} \to \Z^0 \W^{\pm}$ the same way (longitudinal $\W/\Z$'s). \end{subentry} \iteme{PARU(141) :} (D=5.) $\tan\beta$ parameter of a two Higgs doublet scenario, i.e. the ratio of vacuum expectation values. This affects mass relations and couplings in the Higgs sector. \iteme{PARU(142) :} (D=1.) the $\Z^0 \to \H^+ \H^-$ coupling is taken to be {\tt PARU(142)}$\times$(the MSSM $\Z^0 \to \H^+ \H^-$ coupling). \iteme{PARU(143) :} (D=1.) the $\Z'^0 \to \H^+ \H^-$ coupling is taken to be {\tt PARU(143)}$\times$(the MSSM $\Z^0 \to \H^+ \H^-$ coupling). \iteme{PARU(145) :} (D=1.) quadratically multiplicative factor in the $\Z'^0 \to \Z^0 \H^0$ partial width in left--right-symmetric models, expected to be unity (see \cite{Coc91}). \iteme{PARU(146) :} (D=1.) $\sin(2\alpha)$ parameter, enters quadratically as multiplicative factor in the $\W'^{\pm} \to \W^{\pm} \H^0$ partial width in left--right-symmetric models (see \cite{Coc91}). \iteme{PARU(151) :} (D=1.) multiplicative factor in the $\L_{\Q} \to \q \ell$ squared Yukawa coupling, and thereby in the $\L_{\Q}$ partial width and the $\q \ell \to \L_{\Q}$ and other cross sections. Specifically, $\lambda^2/(4\pi) = ${\tt PARU(151)}$\times \alphaem$, i.e. it corresponds to the $k$ factor of \cite{Hew88}. \iteme{PARU(153) :} (D=0.) anomalous magnetic moment of the $\W^{\pm}$; $\eta = \kappa - 1$, where $\eta = 0$ ($\kappa = 1$) is the Standard Model value. \iteme{PARU(155) :} (D=1000. GeV) compositeness scale $\Lambda$. \iteme{PARU(156) :} (D=1.) sign of interference term between standard cross section and composite term ($\eta$ parameter); should be $\pm 1$. \iteme{PARU(157) - PARU(159) :} (D=3*1.) strength of {\bf SU(2)}, {\bf U(1)} and {\bf SU(3)} couplings, respectively, in an excited fermion scenario; cf. $f$, $f'$ and $f_s$ of \cite{Bau90}. \iteme{PARU(161) - PARU(168) :} (D=5*1.,3*0.) multiplicative factors that can be used to modify the default couplings of the $\H^0$ particle in {\Py}. Note that the factors enter quadratically in the partial widths. The default values correspond to the couplings given in the minimal one-Higgs-doublet Standard Model. \begin{subentry} \iteme{PARU(161) :} $\H^0$ coupling to down type quarks. \iteme{PARU(162) :} $\H^0$ coupling to up type quarks. \iteme{PARU(163) :} $\H^0$ coupling to leptons. \iteme{PARU(164) :} $\H^0$ coupling to $\Z^0$. \iteme{PARU(165) :} $\H^0$ coupling to $\W^{\pm}$. \iteme{PARU(168) :} $\H^0$ coupling to $\H^{\pm}$ in $\gamma\gamma \to \H^0$ loops, in MSSM $\sin(\beta-\alpha)+\cos(2\beta)\sin(\beta+\alpha) / (2\scostw)$. \end{subentry} \iteme{PARU(171) - PARU(178) :} (D=7*1.,0.) multiplicative factors that can be used to modify the default couplings of the $\H'^0$ particle in {\Py}. Note that the factors enter quadratically in partial widths. The default values for {\tt PARU(171) - PARU(175)} correspond to the couplings given to $\H^0$ in the minimal one-Higgs-doublet Standard Model, and are therefore not realistic in a two-Higgs-doublet scenario. The default values should be changed appropriately by you. Also the last two default values should be changed; for these the expressions of the minimal supersymmetric Standard Model (MSSM) are given to show parameter normalization. \begin{subentry} \iteme{PARU(171) :} $\H'^0$ coupling to down type quarks. \iteme{PARU(172) :} $\H'^0$ coupling to up type quarks. \iteme{PARU(173) :} $\H'^0$ coupling to leptons. \iteme{PARU(174) :} $\H'^0$ coupling to $\Z^0$. \iteme{PARU(175) :} $\H'^0$ coupling to $W^{\pm}$. \iteme{PARU(176) :} $\H'^0$ coupling to $\H^0 \H^0$, in MSSM $\cos(2\alpha) \cos(\beta+\alpha) - 2 \sin(2\alpha) \sin(\beta+\alpha)$. \iteme{PARU(177) :} $\H'^0$ coupling to $\A^0 \A^0$, in MSSM $\cos(2\beta) \cos(\beta+\alpha)$. \iteme{PARU(178) :} $\H'^0$ coupling to $\H^{\pm}$ in $\gamma \gamma \to \H'^0$ loops, in MSSM $\cos(\beta-\alpha) - \cos(2\beta)\cos(\beta+\alpha) / (2\scostw)$. \end{subentry} \iteme{PARU(181) - PARU(190) :} (D=3*1.,2*0.,2*1.,3*0.) multiplicative factors that can be used to modify the default couplings of the $\A^0$ particle in PYTHIA. Note that the factors enter quadratically in partial widths. The default values for {\tt PARU(181) - PARU(183)} correspond to the couplings given to $\H^0$ in the minimal one-Higgs-doublet Standard Model, and are therefore not realistic in a two-Higgs-doublet scenario. The default values should be changed appropriately by you. {\tt PARU(184)} and {\tt PARU(185)} should be vanishing at the tree level, and are so set; normalization of these couplings agrees with what is used for H and $\H'^0$. Also the other default values should be changed; for these the expressions of the Minimal Supersymmetric Standard Model (MSSM) are given to show parameter normalization. \begin{subentry} \iteme{PARU(181) :} $\A^0$ coupling to down type quarks. \iteme{PARU(182) :} $\A^0$ coupling to up type quarks. \iteme{PARU(183) :} $\A^0$ coupling to leptons. \iteme{PARU(184) :} $\A^0$ coupling to $\Z^0$. \iteme{PARU(185) :} $\A^0$ coupling to $\W^{\pm}$. \iteme{PARU(186) :} $\A^0$ coupling to $\Z^0 \H^0$ (or $\Z^*$ to $\A^0 \H^0$), in MSSM $\cos(\beta-\alpha)$. \iteme{PARU(187) :} $\A^0$ coupling to $\Z^0 \H'^0$ (or $\Z^*$ to $\A^0 \H'^0$), in MSSM $\sin(\beta-\alpha)$. \iteme{PARU(188) :} As {\tt PARU(186)}, but coupling to $\Z'^0$ rather than $\Z^0$. \iteme{PARU(189) :} As {\tt PARU(187)}, but coupling to $\Z'^0$ rather than $\Z^0$. \iteme{PARU(190) :} $\A^0$ coupling to $\H^{\pm}$ in $\gamma \gamma \to \A^0$ loops, 0 in MSSM. \end{subentry} \iteme{PARU(191) - PARU(195) :} (D=4*0.,1.) multiplicative factors that can be used to modify the couplings of the $\H^{\pm}$ particle in {\Py}. Currently only {\tt PARU(195)} is in use. See above for related comments. \begin{subentry} \iteme{PARU(195) :} $\H^{\pm}$ coupling to $\W^{\pm} \H^0$ (or $\W^{* \pm}$ to $\H^{\pm} \H^0$), in MSSM $\cos(\beta-\alpha)$. \end{subentry} \end{entry} \subsection{Further Parameters and Particle Data} \label{ss:parapartdat} The following common blocks are maybe of a more peripheral interest, with the exception of the {\tt MDCY} array, which allows a selective inhibiting of particle decays, and masses of not yet discovered particles, such as {\tt PMAS(6,1)}, the top quark mass. \drawbox{COMMON/LUDAT2/KCHG(500,3),PMAS(500,4),PARF(2000),VCKM(4,4)}% \label{p:LUDAT2} \begin{entry} \itemc{Purpose:} to give access to a number of flavour treatment constants or parameters and particle/parton data. Particle data is stored by compressed code KC rather than by the full KF code. You are reminded that the way to know the KC value is to use the {\tt LUCOMP} function, i.e. {\tt KC = LUCOMP(KF)}. \boxsep \iteme{KCHG(KC,1) :}\label{p:KCHG} three times particle/parton charge for compressed code KC. \iteme{KCHG(KC,2) :} colour information for compressed code KC. \begin{subentry} \iteme{= 0 :} colour-singlet particle. \iteme{= 1 :} quark or antidiquark. \iteme{= -1 :} antiquark or diquark. \iteme{= 2 :} gluon. \end{subentry} \iteme{KCHG(KC,3) :} particle/antiparticle distinction for compressed code KC. \begin{subentry} \iteme{= 0 :} the particle is its own antiparticle. \iteme{= 1 :} a nonidentical antiparticle exists. \end{subentry} \boxsep \iteme{PMAS(KC,1) :}\label{p:PMAS} particle/parton mass $m$ (in GeV) for compressed code KC. \iteme{PMAS(KC,2) :} the total width $\Gamma$ (in GeV) of an assumed symmetric Breit--Wigner mass shape for compressed particle code KC. \iteme{PMAS(KC,3) :} the maximum deviation (in GeV) from the {\tt PMAS(KC,1)} value at which the Breit--Wigner shape above is truncated. (Is used in particle decays, but not in the {\Py} resonance treatment; cf. the {\tt CKIN} variables.) \iteme{PMAS(KC,4) :} the average lifetime $\tau$ for compressed particle code KC, with $c \tau$ in mm, i.e. $\tau$ in units of about $3.33 \times 10^{-12}$ s. \boxsep \iteme{PARF(1) - PARF(60) :}\label{p:PARF} give a parametrization of the $\d\dbar$--$\u\ubar$--$\s\sbar$ flavour mixing in production of flavour-diagonal mesons. Numbers are stored in groups of 10, for the six multiplets pseudoscalar, vector, axial vector ($S=0$), scalar, axial vector ($S=1$) and tensor, in this order; see section \ref{sss:mesonprod}. Within each group, the first two numbers determine the fate of a $\d\dbar$ flavour state, the second two that of a $\u\ubar$ one, the next two that of an $\s\sbar$ one, while the last four are unused. Call the numbers of a pair $p_1$ and $p_2$. Then the probability to produce the state with smallest KF code is $1-p_1$, the probability for the middle one is $p_1 - p_2$ and the probability for the one with largest code is $p_2$, i.e. $p_1$ is the probability to produce either of the two `heavier' ones. \iteme{PARF(61) - PARF(80) :} give flavour {\bf SU(6)} weights for the production of a spin 1/2 or spin 3/2 baryon from a given diquark--quark combination. Should not be changed. \iteme{PARF(101) - PARF(108) :} first five contain $\d$, $\u$, $\s$, $\c$ and $\b$ constituent masses, as to be used in mass formulae, and should not be changed. For $\t$, $\lrm$ and $\hrm$ masses the current values stored in {\tt PMAS(6,1) - PMAS(8,1)} are copied in. \iteme{PARF(111), PARF(112) :} (D=0.0, 0.11 GeV) constant terms in the mass formulae for heavy mesons and baryons, respectively (with diquark getting 2/3 of baryon). \iteme{PARF(113), PARF(114) :} (D=0.16,0.048 GeV) factors which, together with Clebsch-Gordan coefficients and quark constituent masses, determine the mass splitting due to spin-spin interactions for heavy mesons and baryons, respectively. The latter factor is also used for the splitting between spin 0 and spin 1 diquarks. \iteme{PARF(115) - PARF(118) :} (D=0.50, 0.45, 0.55, 0.60 GeV), constant mass terms, added to the constituent masses, to get the mass of heavy mesons with orbital angular momentum $L = 1$. The four numbers are for pseudovector mesons with quark spin 0, and for scalar, pseudovector and tensor mesons with quark spin 1, respectively. \iteme{PARF(121), PARF(122) :} (D=0.1, 0.2 GeV) constant terms, which are subtracted for quark and diquark masses, respectively, in defining the allowed phase space in particle decays into partons. \iteme{PARF(201) - PARF(1960) :} (D=1760*0) relative probabilities for flavour production in the {\tt MSTJ(15)=1} option; to be defined by you before any {\Je} calls. \\ The index in {\tt PARF} is of the compressed form \\ $120 + 80 \times$KTAB1$ + 25 \times$KTABS$ + $KTAB3. \\ Here KTAB1 is the old flavour, fixed by preceding fragmentation history, while KTAB3 is the new flavour, to be selected according to the relevant relative probabilities (except for the very last particle, produced when joining two jets, where both KTAB1 and KTAB3 are known). Only the most frequently appearing quarks/diquarks are defined, according to the code $1 = \d$, $2 = \u$, $3 = \s$, $4 = \c$, $5 = \b$, $6 = \t$, $7 = \d\d_1$, $8 = \u\d_0$, $9 = \u\d_1$, $10 = \u\u_1$, $11 = \s\d_0$, $12 = \s\d_1$, $13 = \s\u_0$, $14 = \s\u_1$, $15 = \s\s_1$, $16 = \c\d_0$, $17 = \c\d_1$, $18 = \c\u_0$, $19 = \c\u_1$, $20 = \c\s_0$, $21 = \c\s_1$, $22 = \c\c_1$. These are thus the only possibilities for the new flavour to be produced; for an occasional old flavour not on this list, the ordinary relative flavour production probabilities will be used. \\ Given the initial and final flavour, the intermediate hadron that is produced is almost fixed. (Initial and final diquark here corresponds to `popcorn' production of mesons intermediate between a baryon and an antibaryon). The additional index KTABS gives the spin type of this hadron, with \\ 0 = pseudoscalar meson or $\Lambda$-like spin 1/2 baryon, \\ 1 = vector meson or $\Sigma$-like spin 1/2 baryon, \\ 2 = tensor meson or spin 3/2 baryon. \\ (Some meson multiplets, not frequently produced, are not accessible by this parametrization.) \\ Note that some combinations of KTAB1, KTAB3 and KTABS do not correspond to a physical particle (a $\Lambda$-like baryon must contain three different quark flavours, a $\Sigma$-like one at least two), and that you must see to it that the corresponding {\tt PARF} entries are vanishing. One additional complication exist when KTAB3 and KTAB1 denote the same flavour content (normally KTAB3$ = $KTAB1, but for diquarks the spin freedom may give KTAB3$ = $KTAB1$\pm 1$): then a flavour neutral meson is to be produced, and here $\d\dbar$, $\u\ubar$ and $\s\sbar$ states mix (heavier flavour states do not, and these are therefore no problem). For these cases the ordinary KTAB3 value gives the total probability to produce either of the mesons possible, while KTAB3$ = $23 gives the relative probability to produce the lightest meson state ($\pi^0$, $\rho^0$, $\a_2^0$), KTAB3$ = $24 relative probability for the middle meson ($\eta$, $\omega$, $\f_2^0$), and KTAB3 = 25 relative probability for the heaviest one ($\eta'$, $\phi$, $f'^0_2$). Note that, for simplicity, these relative probabilities are assumed the same whether initial and final diquark have the same spin or not; the total probability may well be assumed different, however. \\ As a general comment, the sum of {\tt PARF} values for a given KTAB1 need not be normalized to unity, but rather the program will find the sum of relevant weights and normalize to that. The same goes for the KTAB3$ = $23--25 weights. This makes it straightforward to use one common setup of {\tt PARF} values and still switch between different {\tt MSTJ(12)} baryon production modes. \boxsep \iteme{VCKM(I,J) :}\label{p:VCKM} squared matrix elements of the Cabibbo-Kobayashi-Maskawa flavour mixing matrix. \begin{subentry} \iteme{I :} up type generation index, i.e. $1 = \u$, $2 = \c$, $3 = \t$ and $4 = \hrm$. \iteme{J :} down type generation index, i.e. $1 = \d$, $2 = \s$, $3 = \b$ and $4 = \lrm$. \end{subentry} \end{entry} \drawbox{COMMON/LUDAT3/MDCY(500,3),MDME(2000,2),BRAT(2000),% KFDP(2000,5)}\label{p:LUDAT3} \begin{entry} \itemc{Purpose:} to give access to particle decay data and parameters. In particular, the {\tt MDCY(KC,1)} variables may be used to switch on or off the decay of a given particle species, and the {\tt MDME(IDC,1)} ones to switch on or off an individual decay channel of a particle. For quarks, leptons and gauge bosons, a number of decay channels are included that are not allowed for on-mass-shell particles, see {\tt MDME(IDC,2)=102}. These channels are not currently used in {\Je}, but instead find applications in {\Py}. Particle data is stored by compressed code KC rather than by the full KF code. You are reminded that the way to know the KC value is to use the {\tt LUCOMP} function, i.e. {\tt KC = LUCOMP(KF)}. \boxsep \iteme{MDCY(KC,1) :}\label{p:MDCY} switch to tell whether a particle with compressed code KC may be allowed to decay or not. \begin{subentry} \iteme{= 0 :} the particle is not allowed to decay. \iteme{= 1 :} the particle is allowed to decay (if decay information is defined below for the particle). \end{subentry} \iteme{MDCY(KC,2) :} gives the entry point into the decay channel table for compressed particle code KC. Is 0 if no decay channels have been defined. \iteme{MDCY(KC,3) :} gives the total number of decay channels defined for compressed particle code KC, independently of whether they have been assigned a non-vanishing branching ratio or not. Thus the decay channels are found in positions {\tt MDCY(KC,2)} to {\tt MDCY(KC,2)+MDCY(KC,3)-1}. \boxsep \iteme{MDME(IDC,1) :}\label{p:MDME} on/off switch for individual decay channel IDC. In addition, a channel may be left selectively open; this has some special applications in {\Py} which are not currently used in {\Je}. Effective branching ratios are automatically recalculated for the decay channels left open. Also process cross sections are affected; see section \ref{sss:resdecaycross}. If a particle is allowed to decay by the {\tt MDCY(KC,1)} value, at least one channel must be left open by you. A list of decay channels with current IDC numbers may be obtained with {\tt LULIST(12)}. \begin{subentry} \iteme{= -1 :} this is a non-Standard Model decay mode, which by default is assumed not to exist. Normally, this option is used for decays involving fourth generation or $\H^{\pm}$ particles. \iteme{= 0 :} channel is switched off. \iteme{= 1 :} channel is switched on. \iteme{= 2 :} channel is switched on for a particle but off for an antiparticle. It is also on for a particle its own antiparticle, i.e. here it means the same as {\tt =1}. \iteme{= 3 :} channel is switched on for an antiparticle but off for a particle. It is off for a particle its own antiparticle. \iteme{= 4 :} in the production of a pair of equal or charge conjugate resonances in {\Py}, say $\H^0 \to \W^+ \W^-$, either one of the resonances is allowed to decay according to this group of channels, but not both. If the two particles of the pair are different, the channel is on. Within {\Je}, this option only means that the channel is switched off. \iteme{= 5 :} as {\tt =4}, but an independent group of channels, such that in a pair of equal or charge conjugate resonances the decay of either resonance may be specified independently. If the two particles in the pair are different, the channel is off. Within {\Je}, this option only means that the channel is switched off. \itemc{Warning:} the two values -1 and 0 may look similar, but in fact are quite different. In neither case the channel so set is generated, but in the latter case the channel still contributes to the total width of a resonance, and thus affects both simulated line shape and the generated cross section when {\Py} is run. The value 0 is appropriate to a channel we assume exists, even if we are not currently simulating it, while -1 should be used for channels we believe do not exist. In particular, you are warned unwittingly to set fourth generation channels 0 (rather than -1), since by now the support for a fourth generation is small. \itemc{Remark:} all the options above may be freely mixed. The difference, for those cases where both make sense, between using values 2 and 3 and using 4 and 5 is that the latter automatically include charge conjugate states, e.g. $\H^0 \to \W^+ \W^- \to \e^+ \nu_e \d \ubar$ or $\dbar \u \e^- \br{\nu}_e$, but the former only one of them. In calculations of the joint branching ratio, this makes a factor 2 difference. \itemc{Example:} to illustrate the above options, consider the case of a $\W^+ \W^-$ pair. One might then set the following combination of switches for the $\W$:\\ \begin{tabular}{ccl@{\protect\rule{0mm}{\tablinsep}}} channel & value & comment \\ $\u\dbar$ & 1 & allowed for $\W^+$ and $\W^-$ in any combination, \\ $\u\sbar$ & 0 & never produced but contributes to $\W$ width, \\ $\c\dbar$ & 2 & allowed for $\W^+$ only, \\ $\c\sbar$ & 3 & allowed for $\W^-$ only, i.e. properly $\W^- \to \cbar\s$, \\ $\t\bbar$ & 0 & never produced but contributes to $\W$ width \\ & & if the channel is kinematically allowed, \\ $\nu_{\e}\e^+$ & 4 & allowed for one of $\W^+$ or $\W^-$, but not both, \\ $\nu_{\mu}\mu^+$ & 4 & allowed for one of $\W^+$ or $\W^-$, but not both, \\ & & and not in combination with $\nu_{\e}\e^+$, \\ $\nu_{\tau}\tau^+$ & 5 & allowed for the other $\W$, but not both, \\ $\nu_{\chi}\chi^-$ & $-1$ & not produced and does not contribute to $\W$ width.\\ \end{tabular}\\ A $\W^+\W^-$ final state $\u\dbar + \cbar\s$ is allowed, but not its charge conjugate $\ubar\d + \c\sbar$, since the latter decay mode is not allowed for a $\W^+$. The combination $\nu_{\e}\e^+ + \bar{\nu}_{\tau}\tau^-$ is allowed, since the two channels belong to different groups, but not $\nu_{\e}\e^+ + \bar{\nu}_{\mu}\mu^-$, where both belong to the same. Both $\u\dbar + \bar{\nu}_{\tau}\tau^-$ and $\ubar\d + \nu_{\tau}\tau^+$ are allowed, since there is no clash. The full rulebook, for this case, is given by eq. (\ref{eq:WWallchancomb}). A term $r_i^2$ means channel $i$ is allowed for $\W^+$ and $\W^-$ simultaneously, a term $r_i r_j$ that channels $i$ and $j$ may be combined, and a term $2 r_i r_j$ that channels $i$ and $j$ may be combined two ways, i.e. that also a charge conjugate combination is allowed. \end{subentry} \iteme{MDME(IDC,2) :} information on special matrix-element treatment for decay channel IDC. In addition to the outline below, special rules apply for the order in which decay products should be given, so that matrix elements and colour flow is properly treated. One such example is the weak matrix elements, which only will be correct if decay products are given in the right order. The program does not police this, so if you introduce channels of your own and use these codes, you should be guided by the existing particle data. \begin{subentry} \iteme{= 0 :} no special matrix-element treatment; partons and particles are copied directly to the event record, with momentum distributed according to phase space. \iteme{= 1 :} $\omega$ and $\phi$ decays into three pions, eq. (\ref{dec:omegphi}). \iteme{= 2 :} $\pi^0$ or $\eta$ Dalitz decay to $\gamma \e^+ \e^-$, eq. (\ref{dec:Dalitz}). \iteme{= 3 :} used for vector meson decays into two pseudoscalars, to signal non-isotropic decay angle according to eq. (\ref{dec:psvpsps}), where relevant. \iteme{= 4 :} decay of a spin 1 onium resonance to three gluons or to a photon and two gluons, eq. (\ref{ee:Upsilondec}). The gluons may subsequently develop a shower if {\tt MSTJ(23)=1}. \iteme{= 11 :} phase-space production of hadrons from the quarks available. \iteme{= 12 :} as {\tt =11}, but for onia resonances, with the option of modifying the multiplicity distribution separately. \iteme{= 13 :} as {\tt =11}, but at least three hadrons to be produced (useful when the two-body decays are given explicitly). \iteme{= 14 :} as {\tt =11}, but at least four hadrons to be produced. \iteme{= 15 :} as {\tt =11}, but at least five hadrons to be produced. \iteme{= 22 - 30 :} phase-space production of hadrons from the quarks available, with the multiplicity fixed to be {\tt MDME(IDC,2)-20}, i.e. 2--10. \iteme{= 31 :} two or more quarks and particles are distributed according to phase space. If three or more products, the last product is a spectator quark, i.e. sitting at rest with respect to the decaying hadron. \iteme{= 32 :} a $\q\qbar$ or $\g\g$ pair, distributed according to phase space (in angle), and allowed to develop a shower if {\tt MSTJ(23)=1}. \iteme{= 33 :} a triplet $\q X \qbar$, where $X$ is either a gluon or a colour-singlet particle; the final particle ($\qbar$) is assumed to sit at rest with respect to the decaying hadron, and the two first particles ($\q$ and $X$) are allowed to develop a shower if {\tt MSTJ(23)=1}. \iteme{= 41 :} weak decay, where particles are distributed according to phase space, multiplied by a factor from the expected shape of the momentum spectrum of the direct product of the weak decay (the $\nu_{\tau}$ in $\tau$ decay). \iteme{= 42 :} weak decay matrix element for quarks and leptons. Products may be given either in terms of quarks or hadrons, or leptons for some channels. If the spectator system is given in terms of quarks, it is assumed to collapse into one particle from the onset. If the virtual $\W$ decays into quarks, these quarks are converted to particles, according to phase space in the $\W$ rest frame, as in {\tt =11}. Is intended for $\tau$, charm and bottom. \iteme{= 43 :} as {\tt =42}, but if the $\W$ decays into quarks, these will either appear as jets or, for small masses, collapse into a one- or two-body system. \iteme{= 44 :} weak decay matrix element for quarks and leptons, where the spectator system may collapse into one particle for a small invariant mass. If the first two decay products are a $\q\qbar'$ pair, they may develop a parton shower, if {\tt MSTJ(23)=1}. Is intended for top and beyond, but largely superseded by the following option. \iteme{= 45 :} weak decay $\q \to \W \q'$ or $\ell \to \W \nu_{\ell}$, where the $\W$ is registered as a decay product and subsequently treated with {\tt MDME=46}. To distinguish from ordinary $\W$'s on the mass shell, code KF$ = \pm 89$ is used. The virtual $\W$ mass is selected according to the standard weak decay matrix element, times the $\W$ propagator (for {\tt MSTJ(25)=1}). There may be two or three decay products; if a third this is a spectator taken to sit at rest. The spectator system may collapse into one particle. Is intended for top and beyond. \iteme{= 46 :} $\W$ (KF = 89) decay into $\q\qbar'$ or $\ell \nu_{\ell}$ according to relative probabilities given by couplings (as stored in the {\tt BRAT} vector) times a dynamical phase-space factor given by the current $\W$ mass. In the decay, the correct $V-A$ angular distribution is generated if the $\W$ origin is known (heavy quark or lepton). This is therefore the second step of a decay with {\tt MDME=45}. A $\q\qbar'$ pair may subsequently develop a shower. \iteme{= 84 - 88 :} map the decay of this particle onto the generic $\c$, $\b$, $\t$, $\lrm$ or $\hrm$ decay modes defined for KC$ = $84--88. \iteme{= 48 :} as {\tt =42}, but require at least three decay products. \iteme{= 101 :} this is not a proper decay channel, but only to be considered as a continuation line for the decay product listing of the immediately preceding channel. Since the {\tt KFDP} array can contain five decay products per channel, with this code it is possible to define channels with up to ten decay products. It is not allowed to have several continuation lines after each other. \iteme{= 102 :} this is not a proper decay channel for a decaying particle on the mass shell (or nearly so), and is therefore assigned branching ratio 0. For a particle off the mass shell, this decay mode is allowed, however. By including this channel among the others, the switches {\tt MDME(IDC,1)} may be used to allow or forbid these channels in hard processes, with cross sections to be calculated separately. As an example, $\gamma \to \u \ubar$ is not possible for a massless photon, but is an allowed channel in $\ee$ annihilation. \end{subentry} \boxsep \iteme{BRAT(IDC) :}\label{p:BRAT} give branching ratios for the different decay channels. In principle, the sum of branching ratios for a given particle should be unity. Since the program anyway has to calculate the sum of branching ratios left open by the {\tt MDME(IDC,1)} values and normalize to that, you need not explicitly ensure this normalization, however. (Warnings are printed in {\tt LUUPDA(2)} calls if the sum is not unity, but this is entirely intended as a help for finding user mistypings.) For decay channels with {\tt MDME(IDC,2)}$> 80$ the {\tt BRAT} values are dummy. \boxsep \iteme{KFDP(IDC,J) :}\label{p:KFDP} contain the decay products in the different channels, with five positions {\tt J=} 1--5 reserved for each channel IDC. The decay products are given following the standard KF code for jets and particles, with 0 for trailing empty positions. Note that the {\tt MDME(IDC+1,2)=101} option allows you to double the maximum number of decay product in a given channel from 5 to 10, with the five latter products stored {\tt KFDP(IDC+1,J)}. \end{entry} \drawboxtwo{COMMON/LUDAT4/CHAF(500)}{CHARACTER CHAF*8}\label{p:LUDAT4} \begin{entry} \itemc{Purpose:} to give access to character type variables. \boxsep \iteme{CHAF :}\label{p:CHAF} particle names (excluding charge) according to KC code. \end{entry} \subsection{Miscellaneous Comments} The previous sections have dealt with the subroutine options and variables one at a time. This is certainly important, but for a full use of the capabilities of the program, it is also necessary to understand how to make different pieces work together. This is something that cannot be explained fully in a manual, but must also be learnt by trial and error. This section contains some examples of relationships between subroutines, common blocks and parameters. It also contains comments on issues that did not fit in naturally anywhere else, but still might be useful to have on record. \subsubsection{Interfacing to detector simulation} Very often, the output of the program is to be fed into a subsequent detector simulation program. It therefore becomes necessary to set up an interface between the {\tt LUJETS} common block and the detector model. Preferrably this should be done via the {\tt HEPEVT} standard common block, see section \ref{ss:HEPEVT}, but sometimes this may not be convenient. If a {\tt LUEDIT(2)} call is made, the remaining entries exactly correspond to those an ideal detector could see: all non-decayed particles, with the exception of neutrinos. The translation of momenta should be trivial (if need be, a {\tt LUROBO} call can be made to rotate the `preferred' $z$ direction to whatever is the longitudinal direction of the detector), and so should the translation of particle codes. In particular, if the detector simulation program also uses the standard Particle Data Group codes, no conversion at all is needed. The problem then is to select which particles are allowed to decay, and how decay vertex information should be used. Several switches regulate which particles are allowed to decay. First, the master switch {\tt MSTJ(21)} can be used to switch on/off all decays (and it also contains a choice of how fragmentation should be interfaced). Second, a particle must have decay modes defined for it, i.e. the corresponding {\tt MDCY(KC,2)} and {\tt MDCY(KC,3)} entries must be non-zero for compressed code {\tt KC = LUCOMP(KF)}. This is true for all colour neutral particles except the neutrinos, the photon, the proton and the neutron. (This statement is actually not fully correct, since irrelevant `decay modes' with {\tt MDME(IDC,2)=102} exist in some cases.) Third, the individual switch in {\tt MDCY(KC,1)} must be on. Of all the particles with decay modes defined, only $\mu^{\pm}$, $\pi^{\pm}$, $\K^{\pm}$ and $\K_{\mrm{L}}^0$ are by default considered stable. Finally, if {\tt MSTJ(22)} does not have its default value 1, checks are also made on the lifetime of a particle before it is allowed to decay. In the simplest alternative, {\tt MSTJ(22)=2}, the comparison is based on the average lifetime, or rather $c \tau$, measured in mm. Thus if the limit {\tt PARJ(71)} is (the default) 10 mm, then decays of $\K_{\mrm{S}}^0$, $\Lambda$, $\Sigma^-$, $\Sigma^+$, $\Xi^-$, $\Xi^0$ and $\Omega^-$ are all switched off, but charm and bottom still decay. No $c \tau$ values below 1 micron are defined. With the two options {\tt MSTJ(22)=} 3 or 4, a spherical or cylindrical volume is defined around the origin, and all decays taking place inside this volume are ignored. Whenever a particle is in principle allowed to decay, i.e. {\tt MSTJ(21)} and {\tt MDCY} on, an invariant lifetime is selected once and for all and stored in {\tt V(I,5)}. The {\tt K(I,1)} is then also changed to 4. For {\tt MSTJ(22)=1}, such a particle will also decay, but else it could remain in the event record. It is then possible, at a later stage, to expand the volume inside which decays are allowed, and do a new {\tt LUEXEC} call to have particles fulfilling the new conditions (but not the old) decay. As a further option, the {\tt K(I,1)} code may be put to 5, signalling that the particle will definitely decay in the next {\tt LUEXEC} call, at the vertex position given (by the user) in the {\tt V} vector. This then allows the {\Je} decay routines to be used inside a detector simulation program, as follows. For a particle which did not decay before entering the detector, its point of decay is still well defined (in the absence of deflections by electric or magnetic fields), eq. (\ref{dec:newvertex}). If it interacts before that point, the detector simulation program is left to handle things. If not, the {\tt V} vector is updated according to the formula above, {\tt K(I,1)} is set to 5, and {\tt LUEXEC} is called, to give a set of decay products, that can again be tracked. A further possibility is to force particles to decay into specific decay channels; this may be particularly interesting for charm or bottom physics. The choice of channels left open is determined by the values of the switches {\tt MDME(IDC,1)} for decay channel IDC (use {\tt LULIST(12)} to obtain the full listing). One or several channels may be left open; in the latter case effective branching ratios are automatically recalculated without the need for your intervention. It is also possible to differentiate between which channels are left open for particles and which for antiparticles. Lifetimes are not affected by the exclusion of some decay channels. Note that, whereas forced decays can enhance the efficiency for several kinds of studies, it can also introduce unexpected biases, in particular when events may contain several particles with forced decays, cf. section \ref{sss:resdecaycross}. \subsubsection{Parameter values} A non-trivial question is to know which parameter values to use. The default values stored in the program are based on comparisons with LEP $\ee \to \Z^0$ data at around 91 GeV \cite{LEP90}, using a parton-shower picture followed by string fragmentation. If fragmentation is indeed an universal phenomenon, as we would like to think, then the same parameters should also apply at other energies and in other processes. The former aspect, at least, seems to be borne out by comparisons with lower-energy PETRA/PEP data. Note, however, that the choice of parameters is intertwined with the choice of perturbative QCD description. If instead matrix elements are used, a best fit to 30 GeV data would require the values {\tt PARJ(21)=0.40}, {\tt PARJ(41)=1.0} and {\tt PARJ(42)=0.7}. With matrix elements one does not expect an energy independence of the parameters, since the effective minimum invariant mass cut-off is then energy dependent, i.e. so is the amount of soft gluon emission effects lumped together with the fragmentation parameters. This is indeed confirmed by the LEP data. A mismatch in the perturbative QCD treatment could also lead to small differences between different processes. It is often said that the string fragmentation model contains a wealth of parameters. This is certainly true, but it must be remembered that most of these deal with flavour properties, and to a large extent factorize from the treatment of the general event shape. In a fit to the latter it is therefore usually enough to consider the parameters of the perturbative QCD treatment, like $\Lambda$ in $\alphas$ and a shower cut-off $Q_0$ (or $\alphas$ itself and $y_{\mmin}$, if matrix elements are used), the $a$ and $b$ parameter of the Lund symmetric fragmentation function ({\tt PARJ(41)} and {\tt PARJ(42)}) and the width of the transverse momentum distribution ($\sigma = ${\tt PARJ(21)}). In addition, the $a$ and $b$ parameters are very strongly correlated by the requirement of having the correct average multiplicity, such that in a typical $\chi^2$ plot, the allowed region corresponds to a very narrow but very long valley, stretched diagonally from small ($a$,$b$) pairs to large ones. As to the flavour parameters, these are certainly many more, but most of them are understood qualitiatively within one single framework, that of tunnelling pair production of flavours. Since the use of independent fragmentation has fallen in disrespect, it should be pointed out that the default parameters here are not particularly well tuned to the data. This especially applies if one, in addition to asking for independent fragmentation, also asks for another setup of fragmentation functions, i.e. other than the standard Lund symmetric one. In particular, note that most fits to the popular Peterson et al. (SLAC) heavy-flavour fragmentation function are based on the actual observed spectrum. In a Monte Carlo simulation, one must then start out with something harder, to compensate for the energy lost by initial-state photon radiation and gluon bremsstrahlung. Since independent fragmentation is not collinear safe (i.e, the emission of a collinear gluon changes the properties of the final event), the tuning is strongly dependent on the perturbative QCD treatment chosen. All the parameters needed for a tuning of independent fragmentation are available, however. \subsubsection{Particle properties} The masses of most frequently used particles are taken from tables. For some rare charm and bottom hadrons, and for heavier flavour hadrons, this would be unwieldy, and instead mass formulae are used, based on the quark content. For the known quarks $\d$, $\u$, $\s$, $\c$ and $\b$, the masses used for this purpose are actually the ones stored in positions 101--105 in the {\tt PARF} vector, rather than the ones found in {\tt PMAS}. This means that the {\tt PMAS} masses can be freely changed by you, to modify the masses that appear in the event record, without courting disaster elsewhere (since mass formulae typically contain $1/m$ terms from spin-spin splittings, it is necessary to have the non-zero `constituent' masses here). Thus you should never touch the mass values stored in {\tt PARF}. For the heavier flavours top, low and high, the current {\tt PMAS} values are always used. For these flavours, the only individually defined hadrons are the flavour neutral $\eta$, $\Theta$, $\hrm_1$, $\chi_0$, $\chi_1$ and $\chi_2$ states. A complete change of top mass in the program thus requires changing {\tt PMAS(6,1)}, {\tt PMAS(LUCOMP(661),1)}, {\tt PMAS(LUCOMP(663),1)}, {\tt PMAS(LUCOMP(665),1)}, {\tt PMAS(LUCOMP(10661),1)}, {\tt PMAS(LUCOMP(10663),1)} and {\tt PMAS(LUCOMP(20663),1)}. Since the latter heavy-flavour-diagonal states are not normally produced in fragmentation, it would be no disaster to forget changing their masses. Most particles have separately defined decay channels. However, there are so many heavy-flavour hadrons with common decay desciptions, that a few `pseudoparticles' have been introduced for generic decays. The most frequently used ones are 84 for charm decays, 85 for bottom decays and 86 for top decays. Instead of a long list of decay channels, several bottom and charm baryons and all top hadrons therefore only have one `decay channel', which is the instruction to make use of the decay data for particle 84/85/86. The spectator quark of the generic decay channels is found as the light quark/diquark of the hadron considered. All the mesons in the bottom and charm sectors are individually defined, as are the $\Lambda_{\c}$ and $\Lambda_{\b}$ states. For top and heavier hadrons, the decay is likely to be so fast that no hadrons at all are produced, but if they are, the generic pseudoparticle approach is a good approximation. The program contains space so that additional new particles may be introduced. Although not completely trivial, this should not be beyond the ability of an ordinary user. Basically, three steps are involved. First, a mechanism of production has to be introduced. This production may well take place in another program, like {\Py} or some user-written correspondence, where matrix elements are used to select the hard process. In this case the new particle already exists in the {\tt LUJETS} common block when {\Je} is called. A new particle, meson, baryon or glueball, may also be a part of the fragmentation process, in which case {\tt LUKFDI} would have to be suitably modified. The particle might also appear as a decay product from some already existing particle, and then the decay data in {\tt /LUDAT3/} would have to be expanded; conceivably also {\tt LUDECY} would be affected. The second step is to teach to program to recognize the new particle. If a KF code in the range 41 to 80 is used, this is automatically taken care of, and in particular the compressed code KC coincides with KF. If a whole sequence of particles is to be introduced, with KF codes paralleling that of ordinary mesons/baryons (a supersymmetric `meson' multiplet, made of a squark plus an antiquark, say), then {\tt LUCOMP} must be modified to include a mapping from these KF values to currently unused KC ones, like the range 401 - 500. It is the presence of such a mapping that the program uses to accept a given KF code as bona fide. The third and final step is to define the properties of this new particle. Thus particle charge information must be given in {\tt KCHG}, mass, width and lifetime in {\tt PMAS}, particle name in {\tt CHAF}, and decay data in the {\tt MDCY}, {\tt MDME}, {\tt BRAT} and {\tt KFDP} arrays. This process is most conveniently carried out by using {\tt LUUPDA(1)} to produce a table of particle data, which can then be modified, and afterwards read back in with {\tt LUUPDA(2)}. Note that the particle data is to be introduced for the compressed code KC, not for KF proper. \subsection{Examples} A 10 GeV $\u$ quark jet going out along the $+z$ axis is generated with \begin{verbatim} CALL LU1ENT(0,2,10.,0.,0.) \end{verbatim} Note that such a single jet is not required to conserve energy, momentum or flavour. In the generation scheme, particles with negative $p_z$ are produced as well, but these are automatically rejected unless {\tt MSTJ(3)=-1}. While frequently used in former days, the one-jet generation option is not of much current interest. In e.g. a leptoproduction event a typical situation could be a $\u$ quark going out in the $+z$ direction and a $\u\d_0$ target remnant essentially at rest. (Such a process can be simulated by {\Py}, but here we illustrate how to do part of it yourself.) The simplest procedure is probably to treat the process in the c.m. frame and boost it to the lab frame afterwards. Hence, if the c.m. energy is 20 GeV and the boost $\beta_z = 0.996$ (corresponding to $x_B = 0.045$), then \begin{verbatim} CALL LU2ENT(0,2,2101,20.) CALL LUROBO(0.,0.,0.,0.,0.996) \end{verbatim} The jets could of course also be defined and allowed to fragment in the lab frame with \begin{verbatim} CALL LU1ENT(-1,2,223.15,0.,0.) CALL LU1ENT(2,12,0.6837,3.1416,0.) CALL LUEXEC \end{verbatim} Note here that the target diquark is required to move in the backwards direction with $E-p_z = m_{\p}(1-x_B)$ to obtain the correct invariant mass for the system. This is, however, only an artefact of using a fixed diquark mass to represent a varying target remnant mass, and is of no importance for the fragmentation. If one wants a nicer-looking event record, it is possible to use the following \begin{verbatim} CALL LU1ENT(-1,2,223.15,0.,0.) MSTU(10)=1 P(2,5)=0.938*(1.-0.045) CALL LU1ENT(2,2101,0.,0.,0.) MSTU(10)=2 CALL LUEXEC \end{verbatim} A 30 GeV $\u\ubar\g$ event with $E_{\u} = 8$ GeV and $E_{\ubar} = 14$ GeV is simulated with \begin{verbatim} CALL LU3ENT(0,2,21,-2,30.,2.*8./30.,2.*14./30.) \end{verbatim} The event will be given in a standard orientation with the $\u$ quark along the $+z$ axis and the $\ubar$ in the $-z, +x$ quadrant. Note that the flavours of the three partons have to be given in the order they are found along a string, if string fragmentation options are to work. Also note that, for 3-jet events, and particularly 4-jet ones, not all setups of kinematical variables $x$ lie within the kinematically allowed regions of phase space. All common block variables can obviously be changed by including the corresponding common block in the user-written main program. Alternatively, the routine {\tt LUGIVE} can be used to feed in values, with some additional checks on array bounds then performed. A call \begin{verbatim} CALL LUGIVE('MSTJ(21)=3;PMAS(C663,1)=210.;CHAF(401)=funnyino;'// &'PMAS(21,4)=') \end{verbatim} will thus change the value of {\tt MSTJ(21)} to 3, the value of {\tt PMAS(LUCOMP(663),1) = PMAS(136,1)} to 210., the value of {\tt CHAF(401)} to 'funnyino', and print the current value of {\tt PMAS(21,4)}. Since old and new values of parameters changed are written to output, this may offer a convenient way of documenting non-default values used in a given run. On the other hand, if a variable is changed back and forth frequently, the resulting voluminous output may be undesirable, and a direct usage of the common blocks is then to be recommended (the output can also be switched off, see {\tt MSTU(13)}). A general rule of thumb is that none of the physics routines ({\tt LUSTRF}, {\tt LUINDF}, {\tt LUDECY}, etc.) should ever be called directly, but only via {\tt LUEXEC}. This routine may be called repeatedly for one single event. At each call only those entries that are allowed to fragment or decay, and have not yet done so, are treated. Thus \begin{verbatim} CALL LU2ENT(1,1,-1,20.) ! fill 2 jets without fragmenting MSTJ(1)=0 ! inhibit jet fragmentation MSTJ(21)=0 ! inhibit particle decay MDCY(LUCOMP(111),1)=0 ! inhibit pi0 decay CALL LUEXEC ! will not do anything MSTJ(1)=1 ! CALL LUEXEC ! jets will fragment, but no decays MSTJ(21)=2 ! CALL LUEXEC ! particles decay, except pi0 CALL LUEXEC ! nothing new can happen MDCY(LUCOMP(111),1)=1 ! CALL LUEXEC ! pi0:s decay \end{verbatim} A partial exception to the rule above is {\tt LUSHOW}. Its main application is for internal use by {\tt LUEEVT}, {\tt LUDECY}, and {\tt PYEVNT}, but it can also be directly called by you. Note that a special format for storing colour-flow information in {\tt K(I,4)} and {\tt K(I,5)} must then be used. For simple cases, the {\tt LU2ENT} can be made to take care of that automatically, by calling with the first argument negative. \begin{verbatim} CALL LU2ENT(-1,1,-2,80.) ! store d ubar with colour flow CALL LUSHOW(1,2,80.) ! shower partons CALL LUEXEC ! subsequent fragmentation/decay \end{verbatim} For more complicated configurations, {\tt LUJOIN} should be used. It is always good practice to list one or a few events during a run to check that the program is working as intended. With \begin{verbatim} CALL LULIST(1) \end{verbatim} all particles will be listed and in addition total charge, momentum and energy of stable entries will be given. For string fragmentation these quantities should be conserved exactly (up to machine precision errors), and the same goes when running independent fragmentation with one of the momentum conservation options. {\tt LULIST(1)} gives a format that comfortably fits on an 80 column screen, at the price of not giving the complete story. With {\tt LULIST(2)} a more extensive listing is obtained, and {\tt LULIST(3)} also gives vertex information. Further options are available, like {\tt LULIST(12)}, which gives a list of particle data. An event, as stored in the {\tt LUJETS} common block, will contain the original jets and the whole decay chain, i.e. also particles which subsequently decayed. If parton showers are used, the amount of parton information is also considerable: first the on-shell partons before showers have been considered, then a {\tt K(I,1)=22} line with total energy of the showering subsystem, after that the complete shower history tree-like structure, starting off with the same initial partons (now off-shell), and finally the end products of the shower rearranged along the string directions. This detailed record is useful in many connections, but if one only wants to retain the final particles, superfluous information may be removed with {\tt LUEDIT}. Thus e.g. \begin{verbatim} CALL LUEDIT(2) \end{verbatim} will leave you with the final charged and neutral particles, except for neutrinos. The information in {\tt LUJETS} may be used directly to study an event. Some useful additional quantities derived from these, such as charge and rapidity, may easily be found via the {\tt KLU} and {\tt PLU} functions. Thus electric charge {\tt =PLU(I,6)} (as integer, three times charge {\tt =KLU(I,6)}) and true rapidity $y$ with respect to the $z$ axis {\tt = PLU(I,17)}. A number of utility ({\tt MSTU}, {\tt PARU}) and physics ({\tt MSTJ}, {\tt PARJ}) switches and parameters are available in common block {\tt LUDAT1}. All of these have sensible default values. Particle data is stored in common blocks {\tt LUDAT2}, {\tt LUDAT3} and {\tt LUDAT4}. Note that the data in the arrays {\tt KCHG}, {\tt PMAS}, {\tt MDCY} and {\tt CHAF} is not stored by KF code, but by the compressed code KC. This code is not to be learnt by heart, but instead accessed via the conversion function {\tt LUCOMP}, {\tt KC = LUCOMP(KF)}. In the particle tables, the following particles are considered stable: the photon, $\e^{\pm}$, $\mu^{\pm}$, $\pi^{\pm}$, $\K^{\pm}$, $\K_{\mrm{L}}^0$, $\p$, $\pbar$, $\n$, $\br{\n}$ and all the neutrinos. It is, however, always possible to inhibit the decay of any given particle by putting the corresponding {\tt MDCY} value zero or negative, e.g. {\tt MDCY(LUCOMP(310),1)=0} makes $\K_{\mrm{S}}^0$ and {\tt MDCY(LUCOMP(3122),1)=0} $\Lambda$ stable. It is also possible to select stability based on the average lifetime (see {\tt MSTJ(22)}), or based on whether the decay takes place within a given spherical or cylindrical volume around the origin. The Field-Feynman jet model \cite{Fie78} is available in the program by changing the following values: {\tt MSTJ(1)=2} (independent fragmentation), {\tt MSTJ(3)=-1} (retain particles with $p_z < 0$; is not mandatory), {\tt MSTJ(11)=2} (choice of longitudinal fragmentation function, with the $a$ parameter stored in {\tt PARJ(51) - PARJ(53)}), {\tt MSTJ(12)=0} (no baryon production), {\tt MSTJ(13)=1} (give endpoint quarks $\pT$ as quarks created in the field), {\tt MSTJ(24)=0} (no mass broadening of resonances), {\tt PARJ(2)=0.5} ($\s/\u$ ratio for the production of new $\q\qbar$ pairs), {\tt PARJ(11)=PARJ(12)=0.5} (probability for mesons to have spin 1) and {\tt PARJ(21)=0.35} (width of Gaussian transverse momentum distribution). In addition only $\d$, $\u$ and $\s$ single quark jets may be generated following the FF recipe. Today the FF `standard jet' concept is probably dead and buried, so the numbers above should more be taken as an example of the flexibility of the program, than as something to apply in practice. A wide range of independent fragmentation options are implemented, to be accessed with the master switch {\tt MSTJ(1)=2}. In particular, with {\tt MSTJ(2)=1} a gluon jet is assumed to fragment like a random $\d$, $\dbar$, $\u$, $\ubar$, $\s$ or $\sbar$ jet, while with {\tt MSTJ(2)=3} the gluon is split into a $\d\dbar$, $\u\ubar$ or $\s\sbar$ pair of jets sharing the energy according to the Altarelli-Parisi splitting function. Whereas energy, momentum and flavour is not explicitly conserved in independent fragmentation, a number of options are available in {\tt MSTJ(3)} to ensure this `post facto', e.g. {\tt MSTJ(3)=1} will boost the event to ensure momentum conservation and then (in the c.m. frame) rescale momenta by a common factor to obtain energy conservation, whereas {\tt MSTJ(3)=4} rather uses a method of stretching the jets in longitudinal momentum along the respective jet axis to keep angles between jets fixed. \clearpage \section{Event Analysis Routines} To describe the complicated geometries encountered in multihadronic events, a number of event measures have been introduced. These measures are intended to provide a global view of the properties of a given event, wherein the full information content of the event is condensed into one or a few numbers. A steady stream of such measures are proposed for different purposes. Many are rather specialized or never catch on, but a few become standards, and are useful to have easy access to. {\sc Jetset} therefore contains a number of routines that can be called for any event, and that will directly access the event record to extract the required information. In the presentation below, measures have been grouped in three kinds. The first contains simple event shape quantities, such as sphericity and thrust. The second is jet finding algorithms. The third is a mixed bag of particle multiplicities and compositions, factorial moments and energy--energy correlations, put together in a small statistics package. None of the measures presented here are Lorentz invariant. The analysis will be performed in whatever frame the event happens to be given in. It it therefore up to you to decide whether the frame in which events were generated is the right one, or whether events beforehand should be boosted, e.g. to the c.m. frame. You can also decide which particles you want to have affected by the analysis. \subsection{Event Shapes} \label{ss:evshape} In this section we study general event shape variables: sphericity, thrust, Fox-Wolfram moments, and jet masses. These measures are implemented in the routines {\tt LUSPHE}, {\tt LUTHRU}, {\tt LUFOWO} and {\tt LUJMAS}, respectively. Each event is assumed characterized by the particle four-momentum vectors $p_i = (\mbf{p}_i, E_i)$, with $i = 1, 2, \cdots , n$ an index running over the particles of the event. \subsubsection{Sphericity} The sphericity tensor is defined as \cite{Bjo70} \begin{equation} S^{\alpha \beta} = \frac{\displaystyle \sum_i p^{\alpha}_{i} \, p^{\beta}_{i} } {\displaystyle \sum_i |\mbf{p}_i|^2 } ~, \end{equation} where $\alpha, \beta = 1, 2, 3$ corresponds to the $x$, $y$ and $z$ components. By standard diagonalization of $S^{\alpha \beta}$ one may find three eigenvalues $\lambda_1 \geq \lambda_2 \geq \lambda_3$, with $\lambda_1 + \lambda_2 + \lambda_3 = 1$. The sphericity of the event is then defined as \begin{equation} S = \frac{3}{2} \, (\lambda_2 + \lambda_3) ~, \end{equation} so that $0 \leq S \leq 1$. Sphericity is essentially a measure of the summed $\pT^2$ with respect to the event axis; a 2-jet event corresponds to $S \approx 0$ and an isotropic event to $S \approx 1$. The aplanarity $A$, with definition $A = \frac{3}{2} \lambda_3$, is constrained to the range $0 \leq A \leq \frac{1}{2}$. It measures the transverse momentum component out of the event plane: a planar event has $A \approx 0$ and an isotropic one $A \approx \frac{1}{2}$. Eigenvectors $\mbf{v}_j$ can be found that correspond to the three eigenvalues $\lambda_j$ of the sphericity tensor. The $\mbf{v}_1$ one is called the sphericity axis (or event axis, if it is clear from the context that sphericity has been used), while the sphericity event plane is spanned by $\mbf{v}_1$ and $\mbf{v}_2$. The sphericity tensor is quadratic in particle momenta. This means that the sphericity value is changed if one particle is split up into two collinear ones which share the original momentum. Thus sphericity is not an infrared safe quantity in QCD perturbation theory. A useful generalization of the sphericity tensor is \begin{equation} S^{(r) \alpha \beta} = \frac{\displaystyle \sum_i |\mbf{p}_i|^{r-2} \, p^{\alpha}_{i} \, p^{\beta}_{i} }{\displaystyle \sum_i |\mbf{p}_i|^r } ~, \end{equation} where $r$ is the power of the momentum dependence. While $r = 2$ thus corresponds to sphericity, $r = 1$ corresponds to linear measures calculable in perturbation theory \cite{Par78}: \begin{equation} S^{(1) \alpha \beta} = \frac{\displaystyle \sum_i \frac{\displaystyle p^{\alpha}_{i} \, p^{\beta}_{i}}{\displaystyle |\mbf{p}_i|} } {\displaystyle \sum_i |\mbf{p}_i| } ~. \end{equation} Eigenvalues and eigenvectors may be defined exactly as before, and therefore also equivalents of $S$ and $A$. These have no standard names; I tend to call them linearized sphericity $S_{\mrm{lin}}$ and linearized aplanarity $A_{\mrm{lin}}$. Quantities that are standard in the literature are instead the combinations \cite{Ell81} \begin{eqnarray} C & = & 3 ( \lambda_1 \lambda_2 + \lambda_1 \lambda_3 + \lambda_2 \lambda_3 ) ~, \\ D & = & 27 \lambda_1 \lambda_2 \lambda_3 ~. \end{eqnarray} Each of these is constrained to be in the range between 0 and 1. Typically, $C$ is used to measure the 3-jet structure and $D$ the 4-jet one, since $C$ is vanishing for a perfect 2-jet event and $D$ is vanishing for a planar event. The $C$ measure is related to the second Fow-Wolfram moment (see below), $C = 1 - H_2$. Noninteger $r$ values may also be used, and corresponding generalized sphericity and aplanarity measures calculated. While perturbative arguments favour $r = 1$, we know that the fragmentation `noise', e.g. from transverse momentum fluctuations, is proportionately larger for low momentum particles, and so $r > 1$ should be better for experimental event axis determinations. The use of too large an $r$ value, on the other hand, puts all the emphasis on a few high-momentum particles, and therefore involves a loss of information. It should then come as no surprise that intermediate $r$ values, of around 1.5, gives the best performance for event axis determinations in 2-jet events, where the theoretical meaning of the event axis is well-defined. The gain in accuracy compared with the more conventional choices $r=2$ or $r=1$ is rather modest, however. \subsubsection{Thrust} The quantity thrust $T$ is defined by \cite{Bra64} \begin{equation} T = \max_{|\mbf{n}| = 1} \, \frac{\displaystyle \sum_i |\mbf{n} \cdot \mbf{p}_i|} {\displaystyle \sum_i |\mbf{p}_i|} ~, \label{em:thrust} \end{equation} and the thrust axis $\mbf{v}_1$ is given by the $\mbf{n}$ vector for which maximum is attained. The allowed range is $1/2 \leq T \leq 1$, with a 2-jet event corresponding to $T \approx 1$ and an isotropic event to $T \approx 1/2$. In passing, we note that this is not the only definition found in the literature. The definitions agree for events studied in the c.m. frame and where all particles are detected. However, a definition like \begin{equation} T = 2 \, \max_{|\mbf{n}| = 1} \, \frac{\displaystyle \left| \sum_i \theta(\mbf{n} \cdot \mbf{p}_i) \, \mbf{p}_i \right| }{\displaystyle \sum_i |\mbf{p}_i|} = 2 \, \max_{\theta_i = 0,1} \, \frac{\displaystyle \left| \sum_i \theta_i \, \mbf{p}_i \right| } {\displaystyle \sum_i |\mbf{p}_i|} \label{em:thrusttwo} \end{equation} (where $\theta(x)$ is the step function, $\theta(x) = 1 $ if $x > 0$, else $\theta(x) = 0$) gives different results than the one above if e.g. only charged particles are detected. It would even be possible to have $T > 1$; to avoid such problems, often an extra fictitious particle is introduced to balance the total momentum \cite{Bra79}. Eq. (\ref{em:thrust}) may be rewritten as \begin{equation} T = \max_{\epsilon_i = \pm 1} \, \frac{\displaystyle \left| \sum_i \epsilon_i \, \mbf{p}_i \right| } {\displaystyle \sum_i |\mbf{p}_i|} ~. \label{em:thrustthree} \end{equation} (This may also be viewed as applying eq. (\ref{em:thrusttwo}) to an event with $2 n$ particles, $n$ carrying the momenta $\mbf{p}_i$ and $n$ the momenta $- \mbf{p}_i$, thus automatically balancing the momentum.) To find the thrust value and axis this way, $2^{n-1}$ different possibilities would have to be tested. The reduction by a factor of 2 comes from $T$ being unchanged when all $\epsilon_i \to - \epsilon_i$. Therefore this approach rapidly becomes prohibitive. Other exact methods exist, which `only' require about $4n^2$ combinations to be tried. In the implementation in {\Je}, a faster alternative method is used, in which the thrust axis is iterated from a starting direction $\mbf{n}^{(0)}$ according to \begin{equation} \mbf{n}^{(j+1)} = \frac{\displaystyle \sum_i \epsilon( \mbf{n}^{(j)} \cdot \mbf{p}_i) \, \mbf{p}_i } {\displaystyle \left| \sum_i \epsilon( \mbf{n}^{(j)} \cdot \mbf{p}_i) \, \mbf{p}_i \right| } \end{equation} (where $\epsilon(x) = 1$ for $x > 0$ and $\epsilon(x) = -1$ for $x < 0$). It is easy to show that the related thrust value will never decrease, $T^{(j+1)} \geq T^{(j)}$. In fact, the method normally converges in 2--4 iterations. Unfortunately, this convergence need not be towards the correct thrust axis but is occasionally only towards a local maximum of the thrust function \cite{Bra79}. We know of no foolproof way around this complication, but the danger of an error may be lowered if several different starting axes $\mbf{n}^{(0)}$ are tried and found to agree. These $\mbf{n}^{(0)}$ are suitably constructed from the $n'$ (by default 4) particles with the largest momenta in the event, and the $2^{n' -1}$ starting directions $\sum_i \epsilon_i \, \mbf{p}_i$ constructed from these are tried in falling order of the corresponding absolute momentum values. When a predetermined number of the starting axes have given convergence towards the same (best) thrust axis this one is accepted. In the plane perpendicular to the thrust axis, a major \cite{MAR79} axis and value may be defined in just the same fashion as thrust, i.e. \begin{equation} M_a = \max_{|\mbf{n}| = 1, \, \mbf{n} \cdot \mbf{v}_1 = 0} \, \frac{\displaystyle \sum_i |\mbf{n} \cdot \mbf{p}_i|} {\displaystyle \sum_i |\mbf{p}_i| } ~. \end{equation} In a plane more efficient methods can be used to find an axis than in three dimensions \cite{Wu79}, but for simplicity we use the same method as above. Finally, a third axis, the minor axis, is defined perpendicular to the thrust and major ones, and a minor value $M_i$ is calculated just as thrust and major. The difference between major and minor is called oblateness, $O = M_a -M_i$. The upper limit on oblateness depends on the thrust value in a not so simple way. In general $O \approx 0$ corresponds to an event symmetrical around the thrust axis and high $O$ to a planar event. As in the case of sphericity, a generalization to arbitrary momentum dependence may easily be obtained, here by replacing the $\mbf{p}_i$ in the formulae above by $|\mbf{p}_i|^{r-1} \, \mbf{p}_i$. This possibility is included, although so far it has not found any experimental use. \subsubsection{Fox-Wolfram moments} The Fox-Wolfram moments $H_l$, $l = 0, 1, 2, \ldots$, are defined by \cite{Fox79} \begin{equation} H_l = \sum_{i,j} \frac{ |\mbf{p}_i| \, |\mbf{p}_j| } {E_{\mrm{vis}}^2} \, P_l (\cos \theta_{ij}) ~, \end{equation} where $\theta_{ij}$ is the opening angle between hadrons $i$ and $j$ and $E_{\mrm{vis}}$ the total visible energy of the event. Note that also autocorrelations, $i = j$, are included. The $P_l(x)$ are the Legendre polynomials, \begin{eqnarray} P_0(x) & = & 1 ~, \nonumber \\ P_1(x) & = & x ~, \nonumber \\ P_2(x) & = & \frac{1}{2} \, (3x^2 -1) ~, \nonumber \\ P_3(x) & = & \frac{1}{2} \, (5x^3 - 3x) ~, \nonumber \\ P_4(x) & = & \frac{1}{8} \, (35x^4 - 30x^2 + 3) ~. \end{eqnarray} To the extent that particle masses may be neglected, $H_0 \equiv 1$. It is customary to normalize the results to $H_0$, i.e. to give $H_{l0} = H_l / H_0$. If momentum is balanced then $H_1 \equiv 0$. 2-jet events tend to give $H_l \approx 1$ for $l$ even and $\approx 0$ for $l$ odd. \subsubsection{Jet masses} The particles of an event may be divided into two classes. For each class a squared invariant mass may be calculated, $M_1^2$ and $M_2^2$. If the assignment of particles is adjusted such that the sum $M_1^2 + M_2^2$ is minimized, the two masses thus obtained are called heavy and light jet mass, $M_{\mrm{H}}$ and $M_{\mrm{L}}$. It has been shown that these quantities are well behaved in perturbation theory \cite{Cla79}. In $\ee$ annihilation, the heavy jet mass obtains a contribution from $\q\qbar\g$ 3-jet events, whereas the light mass is non-vanishing only when 4-jet events also are included. In the c.m. frame of an event one has the limits $0 \leq M_{\mrm{H}}^2 \leq E_{\mrm{cm}}^2/3$. In general, the subdivision of particles tends to be into two hemispheres, separated by a plane perpendicular to an event axis. As with thrust, it is time-consuming to find the exact solution. Different approximate strategies may therefore be used. In the program, the sphericity axis is used to perform a fast subdivision into two hemispheres, and thus into two preliminary jets. Thereafter one particle at a time is tested to determine whether the sum $M_1^2 + M_2^2$ would be decreased if that particle were to be assigned to the other jet. The precedure is stopped when no further significant change is obtained. Often the original assignment is retained as it is, i.e. the sphericity axis gives a good separation. This is not a full guarantee, since the program might get stuck in a local mimimum which is not the global one. \subsection{Cluster Finding} \label{ss:clustfind} Global event measures, like sphericity or thrust, can only be used to determine the jet axes for back-to-back 2-jet events. To determine the individual jet axes in events with three or more jets, or with two (main) jets which are not back-to-back, cluster algorithms are customarily used. In these, nearby particles are grouped together into a variable number of clusters. Each cluster has a well-defined direction, given by a suitably weighted average of the constituent particle directions. The cluster algorithms traditionally used in $\ee$ and in $\pp$ physics differ in several respects. The former tend to be spherically symmetric, i.e. have no preferred axis in space, and normally all particles have to be assigned to some jet. The latter pick the beam axis as preferred direction, and make use of variables related to this choice, such as rapidity and transverse momentum; additionally only a fraction of all particles are assigned to jets. This reflects a difference in the underlying physics: in $\pp$ collisions, the beam remnants found at low transverse momenta are not related to any hard processes, and therefore only provide an unwanted noise to many studies. (Of course, also hard processes may produce particles at low transverse momenta, but at a rate much less than that from soft or semi-hard processes.) Further, the kinematics of hard processes is, to a good approximation, factorized into the hard subprocess itself, which is boost invariant in rapidity, and parton-distribution effects, which determine the overall position of a hard scattering in rapidity. Hence rapidity, azimuthal angle and transverse momentum is a suitable coordinate frame to describe hard processes in. In standard $\ee$ annihilation events, on the other hand, the hard process c.m. frame tends to be almost at rest, and the event axis is just about randomly distributed in space, i.e. with no preferred r\^ole for the axis defined by the incoming $\e^{\pm}$. All particle production is initiated by and related to the hard subprocess. Some of the particles may be less easy to associate to a specific jet, but there is no compelling reason to remove any of them from consideration. This does not mean that the separation above is always required. $2\gamma$ events in $\ee$ may have a structure with `beam jets' and `hard scattering' jets, for which the $\pp$ type algorithms might be well suited. Conversely, a heavy particle produced in $\pp$ collisions could profitably be studied, in its own rest frame, with $\ee$ techniques. In the following, particles are only characterized by their three-momenta or, alternatively, their energy and direction of motion. No knowledge is therefore assumed of particle types, or even of mass and charge. Clearly, the more is known, the more sophisticated clustering algorithms can be used. The procedure then also becomes more detector-dependent, and therefore less suitable for general usage. {\Je} contains two cluster finding routines. {\tt LUCLUS} is of the $\ee$ type and {\tt LUCELL} of the $\pp$ one. Each of them allows some variations of the basic scheme. \subsubsection{Cluster finding in an $\ee$ type of environment} The usage of cluster algorithms for $\ee$ applications started in the late 1970's. A number of different approaches were proposed \cite{Dor81}. Of these, we will here only discuss those based on binary joining. In this kind of approach, initially each final-state particle is considered to be a cluster. Using some distance measure, the two nearest clusters are found. If their distance is smaller than some cut-off value, the two clusters are joined into one. In this new configuration, the two clusters that are now nearest are found and joined, and so on until all clusters are separated by a distance larger than the cut-off. The clusters remaining at the end are often also called jets. Note that, in this approach, each single particle belongs to exactly one cluster. Also note that the resulting jet picture explicitly depends on the cut-off value used. Normally the number of clusters is allowed to vary from event to event, but occasionally it is more useful to have the cluster algorithm find a predetermined number of jets (like 3). The obvious choice for a distance measure is to use squared invariant mass, i.e. for two clusters $i$ and $j$ to define the distance to be \begin{equation} m^2_{ij} = (E_i + E_j)^2 - (\mbf{p}_i + \mbf{p}_j)^2 ~. \end{equation} (Equivalently, one could have used the invariant mass as measure rather than its square; this is just a matter of convenience.) In fact, a number of people (including the author) tried this measure long ago and gave up on it, since it turns out to have severe instability problems. The reason is well understood: in general, particles tend to cluster closer in invariant mass in the region of small momenta. The clustering process therefore tends to start in the center of the event, and only subsequently spread outwards to encompass also the fast particles. Rather than clustering slow particles around the fast ones (where the latter na\"{\i}vely should best represent the jet directions), the invariant mass measure will tend to cluster fast particles around the slow ones. Another instability may be seen by considering the clustering in a simple 2-jet event. By the time that clustering has reached the level of three clusters, the `best' the clustering algorithm can possibly have achieved, in terms of finding three low-mass clusters, is to have one fast cluster around each jet, plus a third slow cluster in the middle. In the last step this third cluster would be joined with one of the fast ones, to produce two final asymmetric clusters: one cluster would contain all the slow particles, also those that visually look like belonging to the opposite jet. A simple binary joining process, with no possiblity to reassign particles between clusters, is therefore not likely to be optimal. The solution adopted by the author \cite{Sjo83} is to reject invariant mass as distance measure. Instead a jet is defined as a collection of particles which have a limited transverse momentum with respect to a common jet axis, and hence also with respect to each other. This picture is clearly inspired by the standard fragmentation picture, e.g. in string fragmentation. A distance measure $d_{ij}$ between two particles (or clusters) with momenta $\mbf{p}_i$ and $\mbf{p}_j$ should thus not depend critically on the longitudinal momenta but only on the relative transverse momentum. A number of such measures were tried, and the one eventually selected is \begin{equation} d_{ij}^2 = \frac{1}{2} \, (|\mbf{p}_i| \, |\mbf{p}_j| - \mbf{p}_i \cdot \mbf{p}_j) \, \frac{4 \, |\mbf{p}_i| \, |\mbf{p}_j|} {(|\mbf{p}_i| + |\mbf{p}_j|)^2} = \frac{4 \, |\mbf{p}_i|^2 \, |\mbf{p}_j|^2 \, \sin^2(\theta_{ij}/2)} {(|\mbf{p}_i| + |\mbf{p}_j|)^2} ~. \end{equation} For small relative angle $\theta_{ij}$, where $2 \sin(\theta_{ij}/2) \approx \sin\theta_{ij}$ and $\cos \theta_{ij} \approx 1$, this measure reduces to \begin{equation} d_{ij} \approx \frac{|\mbf{p}_i \times \mbf{p}_j|} {|\mbf{p}_i + \mbf{p}_j|} ~, \end{equation} where `$\times$' represents the cross product. We therefore see that $d_{ij}$ in this limit has the simple physical interpretation as the transverse momentum of either particle with respect to the direction given by the sum of the two particle momenta. Unlike the approximate expression, however, $d_{ij}$ does not vanish for two back-to-back particles, but is here more related to the invariant mass between them. The basic scheme is of the binary joining type, i.e. initially each particle is assumed to be a cluster by itself. Then the two clusters with smallest relative distance $d_{ij}$ are found and, if $d_{ij} < d_{\mrm{join}}$, with $d_{\mrm{join}}$ some predetermined distance, the two clusters are joined to one, i.e. their momenta are added vectorially to give the momentum of the new cluster. This is repeated until the distance between any two clusters is $> d_{\mrm{join}}$. The number and momenta of these final clusters then represent our reconstruction of the initial jet configuration, and each particle is assigned to one of the clusters. To make this scheme workable, two further ingredients are necessary, however. Firstly, after two clusters have been joined, some particles belonging to the new cluster may actually be closer to another cluster. Hence, after each joining, all particles in the event are reassigned to the closest of the clusters. For particle $i$, this means that the distance $d_{ij}$ to all clusters $j$ in the event has to be evaluated and compared. After all particles have been considered, and only then, are cluster momenta recalculated to take into account any reassignments. To save time, the assignment procedure is not iterated until a stable configuration is reached, but, since all particles are reassigned at each step, such an iteration is effectively taking place in parallel with the cluster joining. Only at the very end, when all $d_{ij} > d_{\mrm{join}}$, is the reassignment procedure iterated to convergence --- still with the possibility to continue the cluster joining if some $d_{ij}$ should drop below $d_{\mrm{join}}$ due to the reassignment. Occasionally, it may occur that the reassignment step leads to an empty cluster, i.e. one to which no particles are assigned. Since such a cluster has a distance $d_{ij} = 0$ to any other cluster, it is automatically removed in the next cluster joining. However, it is possible to run the program in a mode where a minimum number of jets is to be reconstructed. If this minimum is reached with one cluster empty, the particle is found which has largest distance to the cluster it belongs to. That cluster is then split into two, namely the large-distance particle and a remainder. Thereafter the reassignment procedure is continued as before. Secondly, the large multiplicities normally encountered means that, if each particle initially is to be treated as a separate cluster, the program will become very slow. Therefore a smaller number of clusters, for a normal $\ee$ event typically 8--12, is constructed as a starting point for the iteration above, as follows. The particle with the highest momentum is found, and thereafter all particles within a distance $d_{ij} < d_{\mrm{init}}$ from it, where $d_{\mrm{init}} \ll d_{\mrm{join}}$. Together these are allowed to form a single cluster. For the remaining particles, not assigned to this cluster, the procedure is iterated, until all particles have been used up. Particles in the central momentum region, $|\mbf{p}| < 2d_{\mrm{init}}$ are treated separately; if their vectorial momentum sum is above $2d_{\mrm{init}}$ they are allowed to form one cluster, otherwise they are left unassigned in the initial configuration. The value of $d_{\mrm{init}}$, as long as reasonably small, has no physical importance, in that the same final cluster configuration will be found as if each particle initially is assumed to be a cluster by itself: the particles clustered at this step are so nearby anyway that they almost inevitably must enter the same jet; additionally the reassignment procedure allows any possible `mistake' to be corrected in later steps of the iteration. Thus the jet reconstruction depends on one single parameter, $d_{\mrm{join}}$, with a clearcut physical meaning of a transverse momentum `jet-resolution power'. Neglecting smearing from fragmentation, $d_{ij}$ between two clusters of equal energy corresponds to half the invariant mass of the two original partons. If one only wishes to reconstruct well separated jets, a large $d_{\mrm{join}}$ should be chosen, while a small $d_{\mrm{join}}$ would allow the separation of close jets, at the cost of sometimes artificially dividing a single jet into two. In particular, $\b$ quark jets may here be a nuisance. The value of $d_{\mrm{join}}$ to use for a fixed jet-resolution power in principle should be independent of the c.m. energy of events, although fragmentation effects may give a contamination of spurious extra jets that increases slowly with $E_{\mrm{cm}}$ for fixed $d_{\mrm{join}}$. Therefore a $d_{\mrm{join}} = 2.5$ GeV was acceptable at PETRA/PEP, while 3--4 GeV may be better for applications at LEP and beyond. This completes the description of the main option of the {\tt LUCLUS} routine. Variations are possible. One such is to skip the reassignment step, i.e. to make use only of the simple binary joining procedure, without any possibility to reassign particles between jets. (This option is included mainly as a reference, to check how important reassignment really is.) The other main alternative is to replace the distance measure used above with the one used in the JADE algorithm \cite{JAD86}. The JADE cluster algorithm is an attempt to save the invariant mass measure. The distance measure is defined to be \begin{equation} y_{ij} = \frac{2 E_i E_j (1-\cos \theta_{ij})}{E^2_{\mrm{vis}}} ~. \end{equation} Here $E_{\mrm{vis}}$ is the total visible energy of the event. The usage of $E^2_{\mrm{vis}}$ in the denominator rather than $E_{\mrm{cm}}^2$ tends to make the measure less sensitive to detector acceptance corrections; in addition the dimensionless nature of $y_{ij}$ makes it well suited for a comparison of results at different c.m. energies. For the subsequent discussions, this normalization will be irrelevant, however. The $y_{ij}$ measure is very closely related to the squared mass distance measure: the two coincide (up to the difference in normalization) if $m_i = m_j = 0$. However, consider a pair of particles or clusters with non-vanishing individual masses and a fixed pair mass. Then, the larger the net momentum of the pair, the smaller the $y_{ij}$ measure. This somewhat tends to favour clustering of fast particles, and makes the algorithm less unstable than the one based on true invariant mass. The successes of the JADE algorithm are well known: one obtains a very good agreement between the number of partons generated on the matrix-element (or parton-shower) level and the number of clusters reconstructed from the hadrons, such that QCD aspects like the running of $\alphas$ can be studied with a minimal dependence on fragmentation effects. Of course, the insensitivity to fragmentation effects depends on the choice of fragmentation model. Fragmentation effects are small in the string model, but not in independent fragmentation scenarios. Although independent fragmentation in itself is not credible, this may be seen as a signal for caution. One should note that the JADE measure still suffers from some of the diseases of the simple mass measure (without reassignments), namely that particles which go in opposite directions may well be joined into the same cluster. Therefore, while the JADE algorithm is a good way to find the number of jets, it is inferior to the standard $d_{ij}$ measure for a determination of jet directions and energies \cite{Bet92}. The $d_{ij}$ measure also gives narrower jets, which agree better with the visual impression of jet structure. Recently, the `Durham algorithm' has been introduced \cite{Cat91}, which works as the JADE one but with a distance measure \begin{equation} \tilde{y}_{ij} = \frac{2 \min(E_i^2,E_j^2) (1-\cos \theta_{ij})} {E^2_{cm}} ~. \end{equation} Like the $d_{ij}$ measure, this is a transverse momentum, but $\tilde{y}_{ij}$ has the geometrical interpretation as the transverse momentum of the softer particle with respect to the direction of the harder one, while $d_{ij}$ is the transverse momentum of either particle with respect to the common direction given by the momentum vector sum. The two definitions agree when one cluster is much softer than the other, so the soft gluon exponentiation proven for the Durham measure also holds for the $d_{ij}$ one. The main difference therefore is that the standard {\tt LUCLUS} option allows reassignments, while the Durham algorithm does not. The latter is therefore more easily calculable on the perturbative parton level. This point is sometimes overstressed, and one could give counterexamples why reassignments in fact may bring better agreement with the underlying perturbative level. In particular, without reassignments, one will make the recombination that seems the `best' in the current step, even when that forces you to make `worse' choices in subsequent steps. With reassignments, it is possible to correct for mistakes due to the too local sensitivity of a simple binary joining scheme. \subsubsection{Cluster finding in a $\pp$ type of environment} The {\tt LUCELL} cluster finding routines is of the kind pioneered by UA1 \cite{UA183}, and commonly used in $\pp$ physics. It is based on a choice of pseudorapidity $\eta$, azimuthal angle $\varphi$ and transverse momentum $\pT$ as the fundamental coordinates. This choice is discussed in the introduction to cluster finding above, with the proviso that the theoretically preferred true rapidity has to be replaced by pseudorapidity, to make contact with the real-life detector coordinate system. A fix detector grid is assumed, with the pseudorapidity range $|\eta| < \eta_{\mmax}$ and the full azimuthal range each divided into a number of equally large bins, giving a rectangular grid. The particles of an event impinge on this detector grid. For each cell in ($\eta$,$\varphi$) space, the transverse momentum which enters that cell is summed up to give a total cell $E_{\perp}$ flow. Clearly the model remains very primitive in a number of respects, compared with a real detector. There is no magnetic field allowed for, i.e. also charged particles move in straight tracks. The dimensions of the detector are not specified; hence the positions of the primary vertex and any secondary vertices are neglected when determining which cell a particle belongs to. The rest mass of particles is not taken into account, i.e. what is used is really $\pT = \sqrt{p_x^2 + p_y^2}$, while in a real detector some particles would decay or annihilate, and then deposit additional amounts of energy. To take into account the energy resolution of the detector, it is possible to smear the $E_{\perp}$ contents, bin by bin. This is done according to a Gaussian, with a width assumed proportional to the $\sqrt{E_{\perp}}$ of the bin. The Gaussian is cut off at zero and at some predetermined multiple of the unsmeared $E_{\perp}$, by default twice it. Alternatively, the smearing may be performed in $E$ rather than in $E_{\perp}$. To find the $E$, it is assumed that the full energy of a cell is situated at its center, so that one can translate back and forth with $E = E_{\perp} \cosh\eta_{\mrm{center}}$. The cell with largest $E_{\perp}$ is taken as a jet initiator if its $E_{\perp}$ is above some threshold. A candidate jet is defined to consist of all cells which are within some given radius $R$ in the ($\eta$,$\varphi$) plane, i.e. which have $(\eta - \eta_{\mrm{initiator}})^2 + (\varphi - \varphi_{\mrm{initiator}})^2 < R^2$. Coordinates are always given with respect to the center of the cell. If the summed $E_{\perp}$ of the jet is above the required minimum jet energy, the candidate jet is accepted, and all its cells are removed from further consideration. If not, the candidate is rejected. The sequence is now repeated with the remaining cell of highest $E_{\perp}$, and so on until no single cell fulfills the jet initiator condition. The number of jets reconstruced can thus vary from none to a maximum given by purely geometrical considerations, i.e. how many circles of radius $R$ are needed to cover the allowed ($\eta$,$\varphi$) plane. Normally only a fraction of the particles are assigned to jets. One could consider to iterate the jet assignment process, using the $E_{\perp}$-weighted center of a jet to draw a new cirle of radius $R$. In the current algorithm there is no such iteration step. For an ideal jet assignment it would also be necessary to improve the treatment when two jet circles partially overlap. A final technical note. A natural implementation of a cell finding algorithm is based on having a two-dimensional array of $E_{\perp}$ values, with dimensions to match the detector grid. Very often most of the cells would then be empty, in particular for low-multiplicity events in fine-grained calorimeters. Our implementation is somewhat atypical, since cells are only reserved space (contents and position) when they are shown to be non-empty. This means that all non-empty cells have to be looped over to find which are within the required distance $R$ of a potential jet initiator. The algorithm is therefore faster than the ordinary kind if the average cell occupancy is low, but slower if it is high. \subsection{Event Statistics} All the event-analysis routines above are defined on an event-by-event basis. Once found, the quantities are about equally often used to define inclusive distributions as to select specific classes of events for continued study. For instance, the thrust routine might be used either to find the inclusive $T$ distribution or to select events with $T < 0.9$. Other measures, although still defined for the individual event, only make sense to discuss in terms of averages over many events. A small set of such measures is found in {\tt LUTABU}. This routine has to be called once after each event to accumulate statistics, and once in the end to print the final tables. Of course, among the wealth of possibilities imaginable, the ones collected here are only a small sample, selected because the author at some point has found a use for them himself. \subsubsection{Multiplicities} Three options are available to collect information on multiplicities in events. One gives the flavour content of the final state in hard interaction processes, e.g. the relative composition of $\d\dbar / \u\ubar / \s\sbar / \c\cbar / \b\bbar$ in $\ee$ annihilation events. Additionally it gives the total parton multiplicity distribution at the end of parton showering. Another gives the inclusive rate of all the different particles produced in events, either as intermediate resonances or as final-state particles. The number is subdivided into particles produced from fragmentation (primary particles) and those produced in decays (secondary particles). The third option tabulates the rate of exclusive final states, after all allowed decays have occurred. Since only events with up to 8 final-state particles are analyzed, this is clearly not intended for the study of complete high-energy events. Rather the main application is for an analysis of the decay modes of a single particle. For instance, the decay data for $\D$ mesons is given in terms of channels that also contain unstable particles, such as $\rho$ and $\eta$, which decay further. Therefore a given final state may receive contributions from several tabulated decay channels; e.g. $\K \pi \pi$ from $\K^* \pi$ and $\K \rho$, and so on. \subsubsection{Energy-Energy Correlation} The Energy-Energy Correlation is defined by \cite{Bas78} \begin{equation} \mrm{EEC}(\theta) = \sum_{i < j} \frac{2 E_i E_j}{E_{\mrm{vis}}^2} \, \delta(\theta - \theta_{ij}) ~, \end{equation} and its Asymmetry by \begin{equation} \mrm{EECA}(\theta) = \mrm{EEC}(\pi - \theta) - \mrm{EEC}(\theta) ~. \end{equation} Here $\theta_{ij}$ is the opening angle between the two particles $i$ and $j$, with energies $E_i$ and $E_j$. In principle, normalization should be to $E_{\mrm{cm}}$, but if not all particles are detected it is convenient to normalize to the total visible energy $E_{\mrm{vis}}$. Taking into account the autocorrelation term $i = j$, the total $\mrm{EEC}$ in an event then is unity. The $\delta$ function peak is smeared out by the finite bin width $\Delta \theta$ in the histogram, i.e., it is replaced by a contribution $1 / \Delta \theta$ to the bin which contains $\theta_{ij}$. The formulae above refer to an individual event, and are to be averaged over all events to suppress statistical fluctuations, and obtain smooth functions of $\theta$. \subsubsection{Factorial moments} Factorial moments may be used to search for intermittency in events \cite{Bia86}. The whole field has been much studied in recent years, and a host of different measures have been proposed. We only implement one of the original prescriptions. To calculate the factorial moments, the full rapidity (or pseudorapidity) and azimuthal ranges are subdivided into bins of successively smaller size, and the multiplicity distributions in bins is studied. The program calculates pseudorapidity with respect to the $z$ axis; if desired, one could first find an event axis, e.g. the sphericity or thrust axis, and subsequently rotate the event to align this axis with the $z$ direction. The full rapidity range $|y| < y_{\mmax}$ (or pseudorapidity range $|\eta| < \eta_{\mmax}$) and azimuthal range $0 < \varphi < 2\pi$ are subdivided into $m_y$ and $m_{\varphi}$ equally large bins. In fact, the whole analysis is performed thrice: once with $m_{\varphi}=1$ and the $y$ (or $\eta$) range gradually divided into 1, 2, 4, 8, 16, 32, 64, 128, 256 and 512 bins, once with $m_y = 1$ and the $\varphi$ range subdivided as above, and finally once with $m_y = m_{\varphi}$ according to the same binary sequence. Given the multiplicity $n_j$ in bin $j$, the $i$:th factorial moment is defined by \begin{equation} F_i = (m_y m_{\varphi})^{i-1} \, \sum_j \frac{n_j(n_j-1)\cdots(n_j-i+1)}{n(n-1)\cdots(n-i+1)} ~. \end{equation} Here $n = \sum_j n_j$ is the total multiplicity of the event within the allowed $y$ (or $\eta$) limits. The calculation is performed for the second through the fifth moments, i.e. $F_2$ through $F_5$. The $F_i$ as given here are defined for the individual event, and have to be averaged over many events to give a reasonably smooth behaviour. If particle production is uniform and uncorrelated according to Poissonian statistics, one expects $\langle F_i \rangle \equiv 1$ for all moments and all bin sizes. If, on the other hand, particles are locally clustered, factorial moments should increase when bins are made smaller, down to the characteristic dimensions of the clustering. \subsection{Routines and Common Block Variables} \label{ss:evanrout} The six routines {\tt LUSPHE}, {\tt LUTHRU}, {\tt LUCLUS}, {\tt LUCELL}, {\tt LUJMAS} and {\tt LUFOWO} give you the possibility to find some global event shape properties. The routine {\tt LUTABU} performs a statistical analysis of a number of different quantities like particle content, factorial moments and the energy--energy correlation. Note that, by default, all remaining partons/particles except neutrinos are used in the analysis. Neutrinos may be included with {\tt MSTU(41)=1}. Also note that axes determined are stored in {\tt LUJETS}, but are not proper four-vectors and, as a general rule (with some exceptions), should therefore not be rotated or boosted. \drawbox{CALL LUSPHE(SPH,APL)}\label{p:LUSPHE} \begin{entry} \itemc{Purpose:} to diagonalize the momentum tensor, i.e. find the eigenvalues $\lambda_1 > \lambda_2 > \lambda_3$, with sum unity, and the corresponding eigenvectors. \\ Momentum power dependence is given by {\tt PARU(41)}; default corresponds to sphericity, {\tt PARU(41)=1.} gives measures linear in momenta. Which particles (or partons) are used in the analysis is determined by the {\tt MSTU(41)} value. \iteme{SPH :} $\frac{3}{2} (\lambda_2 + \lambda_3)$, i.e. sphericity (for {\tt PARU(41)=2.}). \begin{subentry} \iteme{= -1. :} analysis not performed because event contained less than two particles (or two exactly back-to-back particles, in which case the two transverse directions would be undefined). \end{subentry} \iteme{APL :} $\frac{3}{2} \lambda_3$, i.e. aplanarity (for {\tt PARU(41)=2.}). \begin{subentry} \iteme{= -1. :} as {\tt SPH=-1.}. \end{subentry} \itemc{Remark:} the lines {\tt N+1} through {\tt N+3} ({\tt N-2} through {\tt N} for {\tt MSTU(43)=2}) in {\tt LUJETS} will, after a call, contain the following information: \\ {\tt K(N+i,1) =} 31; \\ {\tt K(N+i,2) =} 95; \\ {\tt K(N+i,3) :} $i$, the axis number, $i=1,2,3$; \\ {\tt K(N+i,4), K(N+i,5) =} 0; \\ {\tt P(N+i,1) - P(N+i,3) :} the $i$'th eigenvector, $x$, $y$ and $z$ components; \\ {\tt P(N+i,4) :} $\lambda_i$, the $i$'th eigenvalue; \\ {\tt P(N+i,5) =} 0; \\ {\tt V(N+i,1) - V(N+i,5) =} 0. \\ Also, the number of particles used in the analysis is given in {\tt MSTU(62)}. \end{entry} \drawbox{CALL LUTHRU(THR,OBL)}\label{p:LUTHRU} \begin{entry} \itemc{Purpose:} to find the thrust, major and minor axes and corresponding projected momentum quantities, in particular thrust and oblateness. The performance of the program is affected by {\tt MSTU(44)}, {\tt MSTU(45)}, {\tt PARU(42)} and {\tt PARU(48)}. In particular, {\tt PARU(42)} gives the momentum dependence, with the default value {\tt =1.} corresponding to linear dependence. Which particles (or partons) are used in the analysis is determined by the {\tt MSTU(41)} value. \iteme{THR :} thrust (for {\tt PARU(42)=1.}). \begin{subentry} \iteme{= -1. :} analysis not performed because event contained less than two particles. \iteme{= -2. :} remaining space in {\tt LUJETS} (partly used as working area) not large enough to allow analysis. \end{subentry} \iteme{OBL :} oblateness (for {\tt PARU(42)=1.}). \begin{subentry} \iteme{= -1., -2. :} as for {\tt THR}. \end{subentry} \itemc{Remark:} the lines {\tt N+1} through {\tt N+3} ({\tt N-2} through {\tt N} for {\tt MSTU(43)=2}) in {\tt LUJETS} will, after a call, contain the following information: \\ {\tt K(N+i,1) =} 31; \\ {\tt K(N+i,2) =} 96; \\ {\tt K(N+i,3) :} $i$, the axis number, $i=1,2,3$; \\ {\tt K(N+i,4), K(N+i,5) =} 0; \\ {\tt P(N+i,1) - P(N+i,3) :} the thrust, major and minor axis, respectively, for $i = 1, 2$ and 3; \\ {\tt P(N+i,4) :} corresponding thrust, major and minor value; \\ {\tt P(N+i,5) =} 0; \\ {\tt V(N+i,1) - V(N+i,5) =} 0. \\ Also, the number of particles used in the analysis is given in {\tt MSTU(62)}. \end{entry} \drawbox{CALL LUCLUS(NJET)}\label{p:LUCLUS} \begin{entry} \itemc{Purpose:} to reconstruct an arbitrary number of jets using a cluster analysis method based on particle momenta. \\ Two different distance measures are available, see section \ref{ss:clustfind}. The choice is controlled by {\tt MSTU(46)}. The distance scale $d_{\mrm{join}}$, above which two clusters may not be joined, is normally given by {\tt PARU(44)}. In general, $d_{\mrm{join}}$ may be varied to describe different `jet-resolution powers'; the default value, 2.5 GeV, is fairly well suited for $\ee$ physics at 30--40 GeV. With the alternative mass distance measure, {\tt PARU(44)} can be used to set the absolute maximum cluster mass, or {\tt PARU(45)} to set the scaled one, i.e. in $y = m^2/E_{\mrm{cm}}^2$, where $E_{\mrm{cm}}$ is the total invariant mass of the particles being considered. \\ It is possible to continue the cluster search from the configuration already found, with a new higher $d_{\mrm{join}}$ scale, by selecting {\tt MSTU(48)} properly. In {\tt MSTU(47)} one can also require a minimum number of jets to be reconstructed; combined with an artificially large $d_{\mrm{join}}$ this can be used to reconstruct a predetermined number of jets. \\ Which particles (or partons) are used in the analysis is determined by the {\tt MSTU(41)} value, whereas assumptions about particle masses is given by {\tt MSTU(42)}. The parameters {\tt PARU(43)} and {\tt PARU(48)} regulate more technical details (for events at high energies and large multiplicities, however, the choice of a larger {\tt PARU(43)} may be necessary to obtain reasonable reconstruction times). \iteme{NJET :} the number of clusters reconstructed. \begin{subentry} \iteme{= -1 :} analysis not performed because event contained less than {\tt MSTU(47)} (normally 1) particles, or analysis failed to reconstruct the requested number of jets. \iteme{= -2 :} remaining space in {\tt LUJETS} (partly used as working area) not large enough to allow analysis. \end{subentry} \itemc{Remark:} if the analysis does not fail, further information is found in {\tt MSTU(61) - MSTU(63)} and {\tt PARU(61) - PARU(63)}. In particular, {\tt PARU(61)} contains the invariant mass for the system analyzed, i.e. the number used in determining the denominator of $y = m^2/E_{\mrm{cm}}^2$. {\tt PARU(62)} gives the generalized thrust, i.e. the sum of (absolute values of) cluster momenta divided by the sum of particle momenta (roughly the same as multicity). {\tt PARU(63)} gives the minimum distance $d$ (in $\pT$ or $m$) between two clusters in the final cluster configuration, 0 in case of only one cluster. \\ Further, the lines {\tt N+1} through {\tt N+NJET} ({\tt N-NJET+1} through {\tt N} for {\tt MSTU(43)=2}) in {\tt LUJETS} will, after a call, contain the following information: \\ {\tt K(N+i,1) =} 31; \\ {\tt K(N+i,2) =} 97; \\ {\tt K(N+i,3) :} $i$, the jet number, with the jets arranged in falling order of absolute momentum; \\ {\tt K(N+i,4) :} the number of particles assigned to jet $i$; \\ {\tt K(N+i,5) =} 0; \\ {\tt P(N+i,1) - P(N+i,5) :} momentum, energy and invariant mass of jet $i$; \\ {\tt V(N+i,1) - V(N+i,5) =} 0. \\ Also, for a particle which was used in the analysis, {\tt K(I,4)}$ = i$, where {\tt I} is the particle number and $i$ the number of the jet it has ben assigned to. Undecayed particles not used then have {\tt K(I,4)=0}. An exception is made for lines with {\tt K(I,1)=3} (which anyhow are not normally interesting for cluster search), where the colour-flow information stored in {\tt K(I,4)} is left intact. \end{entry} \drawbox{CALL LUCELL(NJET)}\label{p:LUCELL} \begin{entry} \itemc{Purpose:} to provide a simpler cluster routine more in line with what is currently used in the study of high-$\pT$ collider events. \\ A detector is assumed to stretch in pseudorapidity between {\tt -PARU(51)} and {\tt +PARU(51)} and be segmented in {\tt MSTU(51)} equally large $\eta$ (pseudorapidity) bins and {\tt MSTU(52)} $\varphi$ (azimuthal) bins. Transverse energy $E_{\perp}$ for undecayed entries are summed up in each bin. For {\tt MSTU(53)} non-zero, the energy is smeared by calorimetric resolution effects, cell by cell. This is done according to a Gaussian distribution; if {\tt MSTU(53)=1} the standard deviation for the $E_{\perp}$ is {\tt PARU(55)}$\times \sqrt{E_{\perp}}$, if {\tt MSTU(53)=2} the standard deviation for the $E$ is {\tt PARU(55)}$\times \sqrt{E}$, $E_{\perp}$ and $E$ expressed in GeV. The Gaussian is cut off at 0 and at a factor {\tt PARU(56)} times the correct $E_{\perp}$ or $E$. Cells with an $E_{\perp}$ below a given threshold {\tt PARU(58)} are removed from further consideration; by default {\tt PARU(58)=0.} and thus all cells are kept. \\ All bins with $E_{\perp} > ${\tt PARU(52)} are taken to be possible initiators of jets, and are tried in falling $E_{\perp}$ sequence to check whether the total $E_{\perp}$ summed over cells no more distant than {\tt PARU(54)} in $\sqrt{(\Delta\eta)^2 + (\Delta\varphi)^2}$ exceeds {\tt PARU(53)}. If so, these cells define one jet, and are removed from further consideration. Contrary to {\tt LUCLUS}, not all particles need be assigned to jets. Which particles (or partons) are used in the analysis is determined by the {\tt MSTU(41)} value. \iteme{NJET :} the number of jets reconstructed (may be 0). \begin{subentry} \iteme{= -2 :} remaining space in {\tt LUJETS} (partly used as working area) not large enough to allow analysis. \end{subentry} \itemc{Remark:} the lines {\tt N+1} through {\tt N+NJET} ({\tt N-NJET+1} through {\tt N} for {\tt MSTU(43)=2}) in {\tt LUJETS} will, after a call, contain the following information: \\ {\tt K(N+i,1) =} 31; \\ {\tt K(N+i,2) =} 98; \\ {\tt K(N+i,3) :} $i$, the jet number, with the jets arranged in falling order in $E_{\perp}$; \\ {\tt K(N+i,4) :} the number of particles assigned to jet $i$; \\ {\tt K(N+i,5) =} 0; \\ {\tt V(N+i,1) - V(N+i,5) =} 0. \\ Further, for {\tt MSTU(54)=1} \\ {\tt P(N+i,1), P(N+i,2) =} position in $\eta$ and $\varphi$ of the center of the jet initiator cell, i.e. geometrical center of jet; \\ {\tt P(N+i,3), P(N+i,4) =} position in $\eta$ and $\varphi$ of the $E_{\perp}$-weighted center of the jet, i.e. the center of gravity of the jet; \\ {\tt P(N+i,5) =} sum $E_{\perp}$ of the jet; \\ while for {\tt MSTU(54)=2} \\ {\tt P(N+i,1) - P(N+i,5) :} the jet momentum vector, constructed from the summed $E_{\perp}$ and the $\eta$ and $\varphi$ of the $E_{\perp}$-weighted center of the jet as \\ $(p_x, p_y, p_z, E, m)=E_{\perp} (\cos\varphi, \sin\varphi, \sinh\eta, \cosh\eta, 0)$; \\ and for {\tt MSTU(54)=3} \\ {\tt P(N+i,1) - P(N+i,5) :} the jet momentum vector, constructed by adding vectorially the momentum of each cell assigned to the jet, assuming that all the $E_{\perp}$ was deposited at the center of the cell, and with the jet mass in {\tt P(N+i,5)} calculated from the summed $E$ and $\mbf{p}$ as $m^2 = E^2 - p_x^2 - p_y^2 - p_z^2$. \\ Also, the number of particles used in the analysis is given in {\tt MSTU(62)}, and the number of cells hit in {\tt MSTU(63)}. \end{entry} \drawbox{CALL LUJMAS(PMH,PML)}\label{p:LUJMAS} \begin{entry} \itemc{Purpose:} to reconstruct high and low jet mass of an event. A simplified algorithm is used, wherein a preliminary division of the event into two hemispheres is done transversely to the sphericity axis. Then one particle at a time is reassigned to the other hemisphere if that reduces the sum of squares of the two jet masses, $m_{\mrm{H}}^2 + m_{\mrm{L}}^2$. The procedure is stopped when no further significant change (see {\tt PARU(48)}) is obtained. Often, the original assignment is retained as it is. Which particles (or partons) used in the analysis is determined by the {\tt MSTU(41)} value, whereas assumptions about particle masses is given by {\tt MSTU(42)}. \iteme{PMH :} heavy jet mass (in GeV). \begin{subentry} \iteme{= -2. :} remaining space in {\tt LUJETS} (partly used as working area) not large enough to allow analysis. \end{subentry} \iteme{PML :} light jet mass (in GeV). \begin{subentry} \iteme{= -2. :} as for {\tt PMH=-2.}. \end{subentry} \itemc{Remark:} After a successful call, {\tt MSTU(62)} contains the number of particles used in the analysis, and {\tt PARU(61)} the invariant mass of the system analyzed. The latter number is helpful in constructing scaled jet masses. \end{entry} \drawbox{CALL LUFOWO(H10,H20,H30,H40)}\label{p:LUFOWO} \begin{entry} \itemc{Purpose:} to do an event analysis in terms of the Fox-Wolfram moments. The moments $H_i$ are normalized to the lowest one, $H_0$. Which particles (or partons) are used in the analysis is determined by the {\tt MSTU(41)} value. \iteme{H10 :} $H_1/H_0$. Is $=0$ if momentum is balanced. \iteme{H20 :} $H_2/H_0$. \iteme{H30 :} $H_3/H_0$. \iteme{H40 :} $H_4/H_0$. \itemc{Remark:} the number of particles used in the analysis is given in {\tt MSTU(62)}. \end{entry} \drawbox{CALL LUTABU(MTABU)}\label{p:LUTABU} \begin{entry} \itemc{Purpose:} to provide a number of event-analysis options which can be be used on each new event, with accumulated statistics to be written out on request. When errors are quoted, these refer to the uncertainty in the average value for the event sample as a whole, rather than to the spread of the individual events, i.e. errors decrease like one over the square root of the number of events analyzed. For a correct use of {\tt LUTABU}, it is not permissible to freely mix generation and analysis of different classes of events, since only one set of statistics counters exists. A single run may still contain sequential `subruns', between which statistics is reset. Whenever an event is analyzed, the number of particles/partons used is given in {\tt MSTU(62)}. \iteme{MTABU :} determines which action is to be taken. Generally, a last digit equal to 0 indicates that the statistics counters for this option is to be reset; since the counters are reset (by {\tt DATA} statements) at the beginning of a run, this is not used normally. Last digit 1 leads to an analysis of current event with respect to the desired properties. Note that the resulting action may depend on how the event generated has been rotated, boosted or edited before this call. The statistics accumulated is output in tabular form with last digit 2, while it is dumped in the {\tt LUJETS} common block for last digit 3. The latter option may be useful for interfacing to graphics output. \begin{subentry} \iteme{= 10 :} statistics on parton multiplicity is reset. \iteme{= 11 :} the parton content of the current event is analyzed, classified according to the flavour content of the hard interaction and the total number of partons. The flavour content is assumed given in {\tt MSTU(161)} and {\tt MSTU(162)}; these are automatically set e.g. in {\tt LUEEVT} and {\tt PYEVNT} calls. \iteme{= 12 :} gives a table on parton multiplicity distribution. \iteme{= 13 :} stores the parton multiplicity distribution of events in {\tt /LUJETS/}, using the following format: \\ {\tt N =} total number of different channels found; \\ {\tt K(I,1) =} 32; \\ {\tt K(I,2) =} 99; \\ {\tt K(I,3), K(I,4) =} the two flavours of the flavour content; \\ {\tt K(I,5) =} total number of events found with flavour content of {\tt K(I,3)} and {\tt K(I,4)}; \\ {\tt P(I,1) - P(I,5) =} relative probability to find given flavour content and a total of 1, 2, 3, 4 or 5 partons, respectively; \\ {\tt V(I,1) - V(I,5) =} relative probability to find given flavour content and a total of 6--7, 8--10, 11--15, 16--25 or above 25 partons, respectively. \\ In addition, {\tt MSTU(3)=1} and \\ {\tt K(N+1,1) =} 32; \\ {\tt K(N+1,2) =} 99; \\ {\tt K(N+1,5) =} number of events analyzed. \iteme{= 20 :} statistics on particle content is reset. \iteme{= 21 :} the particle/parton content of the current event is analyzed, also for particles which have subsequently decayed and partons which have fragmented (unless this has been made impossible by a preceding {\tt LUEDIT} call). Particles are subdivided into primary and secondary ones, the main principle being that primary particles are those produced in the fragmentation of a string, while secondary come from decay of other particles. Since particles (top, say), may decay into partons, the distinction is not always unique. \iteme{= 22 :} gives a table of particle content in events. \iteme{= 23 :} stores particle content in events in {\tt /LUJETS/}, using the following format: \\ {\tt N =} number of different particle species found; \\ {\tt K(I,1) =} 32; \\ {\tt K(I,2) =} 99; \\ {\tt K(I,3) =} particle KF code; \\ {\tt K(I,5) =} total number of particles and antiparticles of this species; \\ {\tt P(I,1) =} average number of primary particles per event; \\ {\tt P(I,2) =} average number of secondary particles per event; \\ {\tt P(I,3) =} average number of primary antiparticles per event; \\ {\tt P(I,4) =} average number of secondary antiparticles per event; \\ {\tt P(I,5) =} average total number of particles or antiparticles per event. \\ In addition, {\tt MSTU(3)=1} and \\ {\tt K(N+1,1) =} 32; \\ {\tt K(N+1,2) =} 99; \\ {\tt K(N+1,5) =} number of events analyzed; \\ {\tt P(N+1,1) =} average primary multiplicity per event; \\ {\tt P(N+1,2) =} average final multiplicity per event; \\ {\tt P(N+1,3) =} average charged multiplicity per event. \iteme{= 30 :} statistics on factorial moments is reset. \iteme{= 31 :} analyzes the factorial moments of the multiplicity distribution in different bins of rapidity and azimuth. Which particles (or partons) are used in the analysis is determined by the {\tt MSTU(41)} value. The selection between usage of true rapidity, pion rapidity or pseudorapidity is regulated by {\tt MSTU(42)}. The $z$ axis is assumed to be event axis; if this is not desirable find an event axis e.g. with {\tt LUSPHE} or {\tt LUTHRU} and use {\tt LUEDIT(31)}. Maximum (pion-, pseudo-) rapidity, which sets the limit for the rapidity plateau or the experimental acceptance, is given by {\tt PARU(57)}. \iteme{= 32 :} prints a table of the first four factorial moments for various bins of pseudorapidity and azimuth. The moments are properly normalized so that they would be unity (up to statistical fluctuations) for uniform and uncorrelated particle production according to Poissonian statistics, but increasing for decreasing bin size in case of `intermittent' behaviour. The error on the average value is based on the actual statistical sample (i.e. does not use any assumptions on the distribution to relate errors to the average values of higher moments). Note that for small bin sizes, where the average multiplicity is small and the factorial moment therefore only very rarely is non-vanishing, moment values may fluctuate wildly and the errors given may be too low. \iteme{= 33 :} stores the factorial moments in {\tt /LUJETS/}, using the format: \\ {\tt N =} 30, with {\tt I = }$i=1$--10 corresponding to results for slicing the rapidity range in $2^{i-1}$ bins, {\tt I = }$i = 11$--20 to slicing the azimuth in $2^{i-11}$ bins, and {\tt I = }$i = 21$--30 to slicing both rapidity and azimuth, each in $2^{i-21}$ bins; \\ {\tt K(I,1) =} 32; \\ {\tt K(I,2) =} 99; \\ {\tt K(I,3) =} number of bins in rapidity; \\ {\tt K(I,4) =} number of bins in azimuth; \\ {\tt P(I,1) =} rapidity bin size; \\ {\tt P(I,2) - P(I,5) =} $\langle F_2 \rangle$--$\langle F_5 \rangle$, i.e. mean of second, third, fourth and fifth factorial moment; \\ {\tt V(I,1) =} azimuthal bin size; \\ {\tt V(I,2) - V(I,5) =} statistical errors on $\langle F_2 \rangle$--$\langle F_5 \rangle$. \\ In addition, {\tt MSTU(3) =} 1 and \\ {\tt K(31,1) =} 32; \\ {\tt K(31,2) =} 99; \\ {\tt K(31,5) =} number of events analyzed. \iteme{= 40 :} statistics on energy--energy correlation is reset. \iteme{= 41 :} the energy--energy correlation $\mrm{EEC}$ of the current event is analyzed. Which particles (or partons) are used in the analysis is determined by the {\tt MSTU(41)} value. Events are assumed given in their c.m. frame. The weight assigned to a pair $i$ and $j$ is $2 E_i E_j/E_{\mrm{vis}}^2$, where $E_{\mrm{vis}}$ is the sum of energies of all analyzed particles in the event. Energies are determined from the momenta of particles, with mass determined according to the {\tt MSTU(42)} value. Statistics is accumulated for the relative angle $\theta_{ij}$, ranging between 0 and 180 degrees, subdivided into 50 bins. \iteme{= 42 :} prints a table of the energy--energy correlation $\mrm{EEC}$ and its asymmetry $\mrm{EECA}$, with errors. The definition of errors is not unique. In our approach each event is viewed as one observation, i.e. an $\mrm{EEC}$ and $\mrm{EECA}$ distribution is obtained by summing over all particle pairs of an event, and then the average and spread of this event-distribution is calculated in the standard fashion. The quoted error is therefore inversely proportional to the square root of the number of events. It could have been possible to view each single particle pair as one observation, which would have given somewhat lower errors, but then one would also be forced to do a complicated correction procedure to account for the pairs in an event not being uncorrelated (two hard jets separated by a given angle typically corresponds to several pairs at about that angle). Note, however, that in our approach the squared error on an $\mrm{EECA}$ bin is smaller than the sum of the squares of the errors on the corresponding $\mrm{EEC}$ bins (as it should be). Also note that it is not possible to combine the errors of two nearby bins by hand from the information given, since nearby bins are correlated (again a trivial consequence of the presence of jets). \iteme{= 43 :} stores the $\mrm{EEC}$ and $\mrm{EECA}$ in {\tt /LUJETS/}, using the format: \\ {\tt N =} 25; \\ {\tt K(I,1) =} 32; \\ {\tt K(I,2) =} 99; \\ {\tt P(I,1) =} $\mrm{EEC}$ for angles between {\tt I-1} and {\tt I}, in units of $3.6^{\circ}$; \\ {\tt P(I,2) =} $\mrm{EEC}$ for angles between {\tt 50-I} and {\tt 51-I}, in units of $3.6^{\circ}$; \\ {\tt P(I,3) =} $\mrm{EECA}$ for angles between {\tt I-1} and {\tt I}, in units of $3.6^{\circ}$; \\ {\tt P(I,4), P(I,5) :} lower and upper edge of angular range of bin {\tt I}, expressed in radians; \\ {\tt V(I,1) - V(I,3) :} errors on the $\mrm{EEC}$ and $\mrm{EECA}$ values stored in {\tt P(I,1) - P(I,3)} (see {\tt =42} for comments); \\ {\tt V(I,4), V(I,5) :} lower and upper edge of angular range of bin {\tt I}, expressed in degrees. \\ In addition, {\tt MSTU(3)=1} and \\ {\tt K(26,1) =} 32; \\ {\tt K(26,2) =} 99; \\ {\tt K(26,5) =} number of events analyzed. \iteme{= 50 :} statistics on complete final states is reset. \iteme{= 51 :} analyzes the particle content of the final state of the current event record. During the course of the run, statistics is thus accumulated on how often different final states appear. Only final states with up to 8 particles are analyzed, and there is only reserved space for up to 200 different final states. Most high energy events have multiplicities far above 8, so the main use for this tool is to study the effective branching ratios obtained with a given decay model for e.g. charm or bottom hadrons. Then {\tt LU1ENT} may be used to generate one decaying particle at a time, with a subsequent analysis by {\tt LUTABU}. Depending on at what level this studied is to be carried out, some particle decays may be switched off, like $\pi^0$. \iteme{= 52 :} gives a list of the (at most 200) channels with up to 8 particles in the final state, with their relative branching ratio. The ordering is according to multiplicity, and within each multiplicity according to an ascending order of KF codes. The KF codes of the particles belonging to a given channel are given in descending order. \iteme{= 53 :} stores the final states and branching ratios found in {\tt /LUJETS/}, using the format: \\ {\tt N =} number of different explicit final states found (at most 200); \\ {\tt K(I,1) =} 32; \\ {\tt K(I,2) =} 99; \\ {\tt K(I,5) =} multiplicity of given final state, a number between 1 and 8; \\ {\tt P(I,1) - P(I,5), V(I,1) - V(I,3) :} the KF codes of the up to 8 particles of the given final state, converted to real numbers, with trailing zeroes for positions not used; \\ {\tt V(I,5) :} effective branching ratio for the given final state. \\ In addition, {\tt MSTU(3)=1} and \\ {\tt K(N+1,1) =} 32; \\ {\tt K(N+1,2) =} 99; \\ {\tt K(N+1,5) =} number of events analyzed; \\ {\tt V(N+1,5) =} summed branching ratio for finals states not given above, either because they contained more than 8 particles or because all 200 channels have been used up. \end{subentry} \end{entry} \drawbox{COMMON/LUDAT1/MSTU(200),PARU(200),MSTJ(200),PARJ(200)} \begin{entry} \itemc{Purpose:} to give access to a number of status codes and parameters which regulate the performance of {\Je}. Most parameters are described in section \ref{ss:JETswitch}; here only those related to the event-analysis routines are described. \boxsep \iteme{MSTU(41) :}\label{p:MSTU41} (D=2) partons/particles used in the event-analysis routines {\tt LUSPHE}, {\tt LUTHRU}, {\tt LUCLUS}, {\tt LUCELL}, {\tt LUJMAS}, {\tt LUFOWO} and {\tt LUTABU} ({\tt LUTABU(11)} excepted). \begin{subentry} \iteme{= 1 :} all partons/particles that have not fragmented/decayed. \iteme{= 2 :} ditto, with the exception of neutrinos and unknown particles. \iteme{= 3 :} only charged, stable particles, plus any partons still not fragmented. \end{subentry} \iteme{MSTU(42) :} (D=2) assumed particle masses, used in calculating energies $E^2 = \mbf{p}^2 + m^2$, as subsequently used in {\tt LUCLUS}, {\tt LUJMAS} and {\tt LUTABU} (in the latter also for pseudorapidity, pion rapidity or true rapidity selection). \begin{subentry} \iteme{= 0 :} all particles are assumed massless. \iteme{= 1 :} all particles, except the photon, are assumed to have the charged pion mass. \iteme{= 2 :} the true masses are used. \end{subentry} \iteme{MSTU(43) :} (D=1) storing of event-analysis information (mainly jet axes), in {\tt LUSPHE}, {\tt LUTHRU}, {\tt LUCLUS} and {\tt LUCELL}. \begin{subentry} \iteme{= 1 :} stored after the event proper, in positions {\tt N+1} through {\tt N+MSTU(3)}. If several of the routines are used in succession, all but the latest information is overwritten. \iteme{= 2 :} stored with the event proper, i.e. at the end of the event listing, with {\tt N} updated accordingly. If several of the routines are used in succession, all the axes determined are available. \end{subentry} \iteme{MSTU(44) :} (D=4) is the number of the fastest (i.e. with largest momentum) particles used to construct the (at most) 10 most promising starting configurations for the thrust axis determination. \iteme{MSTU(45) :} (D=2) is the number of different starting configurations above, which have to converge to the same (best) value before this is accepted as the correct thrust axis. \iteme{MSTU(46) :} (D=1) distance measure used for the joining of clusters in {\tt LUCLUS}. \begin{subentry} \iteme{= 1 :} $d_{ij}$, i.e. approximately relative transverse momentum. Anytime two clusters have been joined, particles are reassigned to the cluster they now are closest to. The distance cut-off $d_{\mrm{join}}$ is stored in {\tt PARU(44)}. \iteme{= 2 :} distance measure as in {\tt =1}, but particles are never reassigned to new jets. \iteme{= 3 :} JADE distance measyre $y_{ij}$, but with dimensions to correspond approximately to total invariant mass. Particles may never be reassigned between clusters. The distance cut-off $m_{\mmin}$ is stored in {\tt PARU(44)}. \iteme{= 4 :} as {\tt =3}, but a scaled JADE distance $y_{ij}$ is used instead of $m_{ij}$. The distance cut-off $y_{\mmin}$ is stored in {\tt PARU(45)}. \end{subentry} \iteme{MSTU(47) :} (D=1) the minimum number of clusters to be reconstructed by {\tt LUCLUS}. \iteme{MSTU(48) :} (D=0) mode of operation of the {\tt LUCLUS} routine. \begin{subentry} \iteme{= 0 :} the cluster search is started from scratch. \iteme{= 1 :} the clusters obtained in a previous cluster search on the same event (with {\tt MSTU(48)=0}) are to be taken as the starting point for subsequent cluster joining. For this call to have any effect, the joining scale in {\tt PARU(44)} or {\tt PARU(45)} must have been changed. If the event record has been modified after the last {\tt LUCLUS} call, or if any other cluster search parameter setting has been changed, the subsequent result is unpredictable. \end{subentry} \iteme{MSTU(51) :} (D=25) number of pseudorapidity bins that the range between {\tt -PARU(51)} and {\tt +PARU(51)} is divided into to define cell size for {\tt LUCELL}. \iteme{MSTU(52) :} (D=24) number of azimuthal bins, used to define the cell size for {\tt LUCELL}. \iteme{MSTU(53) :} (D=0) smearing of correct energy, imposed cell-by-cell in {\tt LUCELL}, to simulate calorimeter resolution effects. \begin{subentry} \iteme{= 0 :} no smearing. \iteme{= 1 :} the transverse energy in a cell, $E_{\perp}$, is smeared according to a Gaussian distribution with standard deviation {\tt PARU(55)}$\times \sqrt{E_{\perp}}$, where $E_{\perp}$ is given in GeV. The Gaussian is cut off so that $0 < E_{\perp \mrm{smeared}} < ${\tt PARU(56)}$\times E_{\perp \mrm{true}}$. \iteme{= 2 :} as {\tt =1}, but it is the energy $E$ rather than the transverse energy $E_{\perp}$ that is smeared. \end{subentry} \iteme{MSTU(54) :} (D=1) form for presentation of information about reconstructed clusters in {\tt LUCELL}, as stored in {\tt LUJETS} according to the {\tt MSTU(43)} value. \begin{subentry} \iteme{= 1 :} the {\tt P} vector in each line contains $\eta$ and $\varphi$ for the geometric origin of the jet, $\eta$ and $\varphi$ for the weighted center of the jet, and jet $E_{\perp}$, respectively. \iteme{= 2 :} the {\tt P} vector in each line contains a massless four-vector giving the direction of the jet, obtained as \\ $(p_x,p_y,p_z,E,m) = E_{\perp} (\cos\varphi,\sin\varphi,\sinh\eta,\cosh\eta,0)$, \\ where $\eta$ and $\varphi$ give the weighted center of a jet and $E_{\perp}$ its transverse energy. \iteme{= 3 :} the {\tt P} vector in each line contains a massive four-vector, obtained by adding the massless four-vectors of all cells that form part of the jet, and calculating the jet mass from $m^2 = E^2-p_x^2-p_y^2-p_z^2$. For each cell, the total $E_{\perp}$ is summed up, and then translated into a massless four-vector assuming that all the $E_{\perp}$ was deposited in the center of the cell. \end{subentry} \iteme{MSTU(61) :} (I) first entry for storage of event-analysis information in last event analyzed with {\tt LUSPHE}, {\tt LUTHRU}, {\tt LUCLUS} or {\tt LUCELL}. \iteme{MSTU(62) :} (R) number of particles/partons used in the last event analysis with {\tt LUSPHE}, {\tt LUTHRU}, {\tt LUCLUS}, {\tt LUCELL}, {\tt LUJMAS}, {\tt LUFOWO} or {\tt LUTABU}. \iteme{MSTU(63) :} (R) in a {\tt LUCLUS} call, the number of preclusters constructed in order to speed up analysis (should be equal to {\tt MSTU(62)} if {\tt PARU(43)=0.}). In a {\tt LUCELL} call, the number of cells hit. \iteme{MSTU(161), MSTU(162) :}\label{p:MSTU161} hard flavours involved in current event, as used in an analysis with {\tt LUTABU(11)}. Either or both may be set 0, to indicate the presence of one or none hard flavours in event. Is normally set by high-level routines, like {\tt LUEEVT} or {\tt PYEVNT}, but can also be set by you. \boxsep \iteme{PARU(41) :}\label{p:PARU41} (D=2.) power of momentum-dependence in {\tt LUSPHE}, default corresponds to sphericity, {\tt =1.} to linear event measures. \iteme{PARU(42) :} (D=1.) power of momentum-dependence in {\tt LUTHRU}, default corresponds to thrust. \iteme{PARU(43) :} (D=0.25 GeV) maximum distance $d_{\mrm{init}}$ allowed in {\tt LUCLUS} when forming starting clusters used to speed up reconstruction. The meaning of the parameter is in $\pT$ for {\tt MSTU(46)}$\leq 2$ and in $m$ for {\tt MSTU(46)}$\geq 3$. If {\tt =0.}, no preclustering is obtained. If chosen too large, more joining may be generated at this stage than is desirable. The main application is at high energies, where some speedup is imperative, and the small details are not so important anyway. \iteme{PARU(44) :} (D=2.5 GeV) maximum distance $d_{\mrm{join}}$, below which it is allowed to join two clusters into one in {\tt LUCLUS}. Is used for {\tt MSTU(46)}$\leq 3$, i.e. both for $\pT$ and mass distance measure. \iteme{PARU(45) :} (D=0.05) maximum distance $y_{\mrm{join}} = m^2/E_{\mrm{vis}}^2$, below which it is allowed to join two clusters into one in {\tt LUCLUS} for {\tt MSTU(46)=4}. \iteme{PARU(48) :} (D=0.0001) convergence criterion for thrust (in {\tt LUTHRU}) or generalized thrust (in {\tt LUCLUS}), or relative change of $m_{\mrm{H}}^2 + m_{\mrm{L}}^2$ (in {\tt LUJMAS}), i.e. when the value changes by less than this amount between two iterations the process is stopped. \iteme{PARU(51) :} (D=2.5) defines maximum absolute pseudorapidity used for detector assumed in {\tt LUCELL}. \iteme{PARU(52) :} (D=1.5 GeV) gives minimum $E_{\perp}$ for a cell to be considered as a potential jet initiator by {\tt LUCELL}. \iteme{PARU(53) :} (D=7.0 GeV) gives minimum summed $E_{\perp}$ for a collection of cells to be accepted as a jet. \iteme{PARU(54) :} (D=1.) gives the maximum distance in $R = \sqrt{(\Delta\eta)^2 + (\Delta\varphi)^2}$ from cell initiator when grouping cells to check whether they qualify as a jet. \iteme{PARU(55) :} (D=0.5) when smearing the transverse energy (or energy, see {\tt MSTU(53)}) in {\tt LUCELL}, the calorimeter cell resolution is taken to be {\tt PARU(55)}$\times \sqrt{E_{\perp}}$ (or {\tt PARU(55)}$\times \sqrt{E}$) for $E_{\perp}$ (or $E$) in GeV. \iteme{PARU(56) :} (D=2.) maximum factor of upward fluctuation in transverse energy or energy in a given cell when calorimeter resolution is included in {\tt LUCELL} (see {\tt MSTU(53)}). \iteme{PARU(57) :} (D=3.2) maximum rapidity (or pseudorapidity or pion rapidity, depending on {\tt MSTU(42)}) used in the factorial moments analysis in {\tt LUTABU}. \iteme{PARU(58) :} (D=0. GeV) in a {\tt LUCELL} call, cells with a transverse energy $E_{\perp}$ below {\tt PARP(58)} are removed from further consideration. This may be used to represent a threshold in an actual calorimeter, or may be chosen just to speed up the algorithm in a high-multiplicity environment.. \iteme{PARU(61) :} (I) invariant mass $W$ of a system analyzed with {\tt LUCLUS} or {\tt LUJMAS}, with energies calculated according to the {\tt MSTU(42)} value. \iteme{PARU(62) :} (R) the generalized thrust obtained after a successful {\tt LUCLUS} call, i.e. ratio of summed cluster momenta and summed particle momenta. \iteme{PARU(63) :} (R) the minimum distance $d$ between two clusters in the final cluster configuration after a successful {\tt LUCLUS} call; is 0 if only one cluster left. \end{entry} \clearpage \section{Summary and Outlook} A complete description of the {\PyJe} programs would have to cover four aspects: \begin{Enumerate} \item the basic philosophy and principles underlying the programs; \item the detailed physics scenarios implemented, with all the necessary compromises and approximations; \item the structure of the implementation, including program flow, internal variable names and programming tricks; and \item the manual, which describes how to use the programs. \end{Enumerate} Of these aspects, the first has been dealt with in reasonable detail. The second is unevenly covered: in depth for aspects which are not discussed anywhere else, more summarily for areas where separate up-to-date papers already exist. The third is not included at all, but `left as an exercise' for the reader, to figure out from the code itself. The fourth, finally, should be largely covered, although many further comments could have been made, in particular about the interplay between different parts of the programs. Still, in the end, no manual, however complete, can substitute for `hands on' experience. The {\PyJe} programs are continuously being developed. We are aware of many shortcomings, which hopefully will be addressed in the future, such as: \begin{Itemize} \item polarization effects should be included in more places, in particular for $\tau$ production and decay; \item the photoproduction and $\gamma\gamma$-physics scenarios should be expanded; \item many processes of interest are missing; and \item mass relations and couplings need to be included beyond the Born level in the MSSM two Higgs doublet scenario. \end{Itemize} This list could have been made much longer (I almost certainly missed your top priority). One other aspect would be to provide more and longer examples of working main programs for a number of standard applications. Apart from these physics aspects, one may also worry about the programming ones. For instance, for historical reasons, single precision real is used almost everywhere. With the push to higher energies, this is becoming more and more of a problem, so it would be logical to move to double precision throughout. One should also note that the {\Je} and {\Py} programs these days are becoming so intertwined, that it would make sense to join them into one single program. This would e.g. mean that the current $\ee$ generation routines of {\Je} are made part of the generic {\Py} process generation machinery --- this is particular affects the matrix-element options, since $\ee$ events with parton showers already are available in {\Py}. A joint product would likely adopt the name {\Py}: although {\Je} is the older of the two programs, it has a less well developed identity of its own. (It is also often referred to as `Lund', which today is more confusing than it was in the early days.) In the process of joining the programs, one would probably also remove a number of options that are no longer used. Another possible change on longer time scales would be an introduction of Fortran 90 programming elements. In particular, derived data types could be used to define the event record as a one-dimensional array, where each element represents a particle, with integer and real components to give flavour, history, momentum and production vertex. No timetable is set up for future changes. After all, this is not a professionally maintained software product, but part of a one-man physics research project. Very often, developments of the programs have come about as a direct response to the evolution of the physics stage, i.e. experimental results and studies for future accelerators. Hopefully, the program will keep on evolving in step with the new challenges opening up. \clearpage \begin{thebibliography}{MMM99x} \addcontentsline{toc}{section}{References} \bibitem[Abb87]{Abb87} A. Abbasabadi and W. Repko, Phys. Lett. {\bf B199} (1987) 286; Phys. Rev. {\bf D37} (1988) 2668; \\ W. Repko and G.L. 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Rev. {\bf D35} (1987) 3396 \bibitem[Zer90]{Zer90} P. Zerwas, private communication \bibitem[Zhu83]{Zhu83} R.-y. Zhu, Ph. D. Thesis (M.I.T.), MIT-LNS Report RX-1033 (1983); Caltech Report CALT-68-1306; in Proceedings of the 1984 DPF conference, Santa Fe, p. 229; in Proceedings of 1985 DPF conference, Oregon, p. 552 \end{thebibliography} \clearpage \section*{Index of Subprograms and Common Block Variables} \addcontentsline{toc}{section}% {Index of Subprograms and Common Block Variables} This index is not intended to be complete, but gives the page where the main description begins of a subroutine, function, block data, common block, variable or array. For common block variables also the name of the common block is given. When some components of an array are described in a separate place, a special reference (indented with respect to the main one) is given for these components. \boxsep \noindent \begin{minipage}[t]{\halfpagewid} \begin{tabular*}{\halfpagewid}[t]{@{}l@{\extracolsep{\fill}}r@{}} {\tt BRAT} in {\tt LUDAT3} & \pageref{p:BRAT} \\ {\tt CHAF} in {\tt LUDAT4} & \pageref{p:CHAF} \\ {\tt CKIN} in {\tt PYSUBS} & \pageref{p:CKIN} \\ {\tt COEF} in {\tt PYINT2} & \pageref{p:COEF} \\ {\tt HEPEVT} common block & \pageref{p:HEPEVT} \\ {\tt ICOL} in {\tt PYINT2} & \pageref{p:ICOL} \\ {\tt IFUP} in {\tt PYUPPR} & \pageref{p:IFUP} \\ {\tt ISET} in {\tt PYINT2} & \pageref{p:ISET} \\ {\tt ISIG} in {\tt PYINT3} & \pageref{p:ISIG} \\ {\tt K} in {\tt LUJETS} & \pageref{p:K} \\ {\tt KCHG} in {\tt LUDAT2} & \pageref{p:KCHG} \\ {\tt KFDP} in {\tt LUDAT3} & \pageref{p:KFDP} \\ {\tt KFIN} in {\tt PYSUBS} & \pageref{p:KFIN} \\ {\tt KFPR} in {\tt PYINT2} & \pageref{p:KFPR} \\ {\tt KLU} function & \pageref{p:KLU} \\ {\tt KUP} in {\tt PYUPPR} & \pageref{p:KUP} \\ {\tt LU1ENT} subroutine & \pageref{p:LU1ENT} \\ {\tt LU2ENT} subroutine & \pageref{p:LU2ENT} \\ {\tt LU3ENT} subroutine & \pageref{p:LU3ENT} \\ {\tt LU4ENT} subroutine & \pageref{p:LU4ENT} \\ {\tt LUBOEI} subroutine & \pageref{p:LUBOEI} \\ {\tt LUCELL} subroutine & \pageref{p:LUCELL} \\ {\tt LUCHGE} function & \pageref{p:LUCHGE} \\ {\tt LUCLUS} subroutine & \pageref{p:LUCLUS} \\ {\tt LUCOMP} function & \pageref{p:LUCOMP} \\ {\tt LUDAT1} common block & \pageref{p:LUDAT1} \\ {\tt LUDAT2} common block & \pageref{p:LUDAT2} \\ {\tt LUDAT3} common block & \pageref{p:LUDAT3} \\ {\tt LUDAT4} common block & \pageref{p:LUDAT4} \\ {\tt LUDATA} block data & \pageref{p:LUDATA} \\ {\tt LUDATR} common block & \pageref{p:LUDATR} \\ {\tt LUDBRB} subroutine & \pageref{p:LUDBRB} \\ {\tt LUDECY} subroutine & \pageref{p:LUDECY} \\ {\tt LUEDIT} subroutine & \pageref{p:LUEDIT} \\ {\tt LUEEVT} subroutine & \pageref{p:LUEEVT} \\ {\tt LUERRM} subroutine & \pageref{p:LUERRM} \\ {\tt LUEXEC} subroutine & \pageref{p:LUEXEC} \\ {\tt LUFOWO} subroutine & \pageref{p:LUFOWO} \\ {\tt LUGIVE} subroutine & \pageref{p:LUGIVE} \\ {\tt LUHEPC} subroutine & \pageref{p:LUHEPC} \\ {\tt LUINDF} subroutine & \pageref{p:LUINDF} \\ {\tt LUJETS} common block & \pageref{p:LUJETS} \\ \end{tabular*} \end{minipage}% \hfill% \begin{minipage}[t]{\halfpagewid} \begin{tabular*}{\halfpagewid}[t]{@{}l@{\extracolsep{\fill}}r@{}} {\tt LUJMAS} subroutine & \pageref{p:LUJMAS} \\ {\tt LUJOIN} subroutine & \pageref{p:LUJOIN} \\ {\tt LUKFDI} subroutine & \pageref{p:LUKFDI} \\ {\tt LULIST} subroutine & \pageref{p:LULIST} \\ {\tt LULOGO} subroutine & \pageref{p:LULOGO} \\ {\tt LUNAME} subroutine & \pageref{p:LUNAME} \\ {\tt LUONIA} subroutine & \pageref{p:LUONIA} \\ {\tt LUPREP} subroutine & \pageref{p:LUPREP} \\ {\tt LUPTDI} subroutine & \pageref{p:LUPTDI} \\ {\tt LURADK} subroutine & \pageref{p:LURADK} \\ {\tt LUROBO} subroutine & \pageref{p:LUROBO} \\ {\tt LUSHOW} subroutine & \pageref{p:LUSHOW} \\ {\tt LUSPHE} subroutine & \pageref{p:LUSPHE} \\ {\tt LUSTRF} subroutine & \pageref{p:LUSTRF} \\ {\tt LUTABU} subroutine & \pageref{p:LUTABU} \\ {\tt LUTEST} subroutine & \pageref{p:LUTEST} \\ {\tt LUTHRU} subroutine & \pageref{p:LUTHRU} \\ {\tt LUUPDA} subroutine & \pageref{p:LUUPDA} \\ {\tt LUX3JT} subroutine & \pageref{p:LUX3JT} \\ {\tt LUX4JT} subroutine & \pageref{p:LUX4JT} \\ {\tt LUXDIF} subroutine & \pageref{p:LUXDIF} \\ {\tt LUXJET} subroutine & \pageref{p:LUXJET} \\ {\tt LUXKFL} subroutine & \pageref{p:LUXKFL} \\ {\tt LUXTOT} subroutine & \pageref{p:LUXTOT} \\ {\tt LUZDIS} subroutine & \pageref{p:LUZDIS} \\ {\tt MDCY} in {\tt LUDAT3} & \pageref{p:MDCY} \\ {\tt MDME} in {\tt LUDAT3} & \pageref{p:MDME} \\ {\tt MINT} in {\tt PYINT1} & \pageref{p:MINT} \\ {\tt MRLU} in {\tt LUDATR} & \pageref{p:MRLU} \\ {\tt MSEL} in {\tt PYSUBS} & \pageref{p:MSEL} \\ {\tt MSTI} in {\tt PYPARS} & \pageref{p:MSTI} \\ {\tt MSTJ} in {\tt LUDAT1}, main & \pageref{p:MSTJ} \\ ~~~{\tt MSTJ(41) - MSTJ(50)} & \pageref{p:MSTJ41} \\ ~~~{\tt MSTJ(101) - MSTJ(121)} & \pageref{p:MSTJ101} \\ {\tt MSTP} in {\tt PYPARS}, main & \pageref{p:MSTP} \\ ~~~{\tt MSTP(22)} & \pageref{p:MSTP22} \\ ~~~{\tt MSTP(61) - MSTP(71)} & \pageref{p:MSTP61} \\ ~~~{\tt MSTP(81) - MSTP(93)} & \pageref{p:MSTP81} \\ ~~~{\tt MSTP(131) - MSTP(134)} & \pageref{p:MSTP131} \\ {\tt MSTU} in {\tt LUDAT1}, main & \pageref{p:MSTU} \\ ~~~{\tt MSTU(41) - MSTU(63)} & \pageref{p:MSTU41} \\ ~~~{\tt MSTU(101) - MSTU(118)} & \pageref{p:MSTU101} \\ \end{tabular*} \end{minipage} \newpage \noindent \begin{minipage}[t]{\halfpagewid} \begin{tabular*}{\halfpagewid}[t]{@{}l@{\extracolsep{\fill}}r@{}} ~~~{\tt MSTU(161) - MSTU(162)} & \pageref{p:MSTU161} \\ {\tt MSUB} in {\tt PYSUBS} & \pageref{p:MSUB} \\ {\tt N} in {\tt LUJETS} & \pageref{p:N} \\ {\tt NFUP} in {\tt PYUPPR} & \pageref{p:NFUP} \\ {\tt NGEN} in {\tt PYINT5} & \pageref{p:NGEN} \\ {\tt NUP} in {\tt PYUPPR} & \pageref{p:NUP} \\ {\tt P} in {\tt LUJETS} & \pageref{p:P} \\ {\tt PARF} in {\tt LUDAT2} & \pageref{p:PARF} \\ {\tt PARI} in {\tt PYPARS} & \pageref{p:PARI} \\ {\tt PARJ} in {\tt LUDAT1}, main & \pageref{p:PARJ} \\ ~~~{\tt PARJ(81) - PARJ(88)} & \pageref{p:PARJ81} \\ ~~~{\tt PARJ(121) - PARJ(171)} & \pageref{p:PARJ121} \\ {\tt PARP} in {\tt PYPARS}, main & \pageref{p:PARP} \\ ~~~{\tt PARP(61) - PARP(71)} & \pageref{p:PARP61} \\ ~~~{\tt PARP(81) - PARP(100)} & \pageref{p:PARP81} \\ ~~~{\tt PARP(131)} & \pageref{p:PARP131} \\ {\tt PARU} in {\tt LUDAT1}, main & \pageref{p:PARU} \\ ~~~{\tt PARU(41) - PARU(63)} & \pageref{p:PARU41} \\ ~~~{\tt PARU(101) - PARU(195)} & \pageref{p:PARU101} \\ {\tt PLU} function & \pageref{p:PLU} \\ {\tt PMAS} in {\tt LUDAT2} & \pageref{p:PMAS} \\ {\tt PROC} in {\tt PYINT6} & \pageref{p:PROC} \\ {\tt PUP} in {\tt PYUPPR} & \pageref{p:PUP} \\ {\tt PYDATA} block data & \pageref{p:PYDATA} \\ {\tt PYDIFF} subroutine & \pageref{p:PYDIFF} \\ {\tt PYDILN} subroutine & \pageref{p:PYDILN} \\ {\tt PYDOCU} subroutine & \pageref{p:PYDOCU} \\ {\tt PYEVKI} subroutine & \pageref{p:PYEVKI} \\ {\tt PYEVNT} subroutine & \pageref{p:PYEVNT} \\ {\tt PYEVWT} subroutine & \pageref{p:PYEVWT} \\ {\tt PYFRAM} subroutine & \pageref{p:PYFRAM} \\ {\tt PYGAMM} function & \pageref{p:PYGAMM} \\ {\tt PYHFTH} function & \pageref{p:PYHFTH} \\ {\tt PYI3AU} subroutine & \pageref{p:PYI3AU} \\ {\tt PYINIT} subroutine & \pageref{p:PYINIT} \\ {\tt PYINKI} subroutine & \pageref{p:PYINKI} \\ {\tt PYINRE} subroutine & \pageref{p:PYINRE} \\ {\tt PYINT1} common block & \pageref{p:PYINT1} \\ {\tt PYINT2} common block & \pageref{p:PYINT2} \\ {\tt PYINT3} common block & \pageref{p:PYINT3} \\ {\tt PYINT4} common block & \pageref{p:PYINT4} \\ {\tt PYINT5} common block & \pageref{p:PYINT5} \\ {\tt PYINT6} common block & \pageref{p:PYINT6} \\ {\tt PYINT7} common block & \pageref{p:PYINT7} \\ {\tt PYKCUT} subroutine & \pageref{p:PYKCUT} \\ {\tt PYKLIM} subroutine & \pageref{p:PYKLIM} \\ {\tt PYKMAP} subroutine & \pageref{p:PYKMAP} \\ {\tt PYMAXI} subroutine & \pageref{p:PYMAXI} \\ {\tt PYMULT} subroutine & \pageref{p:PYMULT} \\ \end{tabular*} \end{minipage}% \hfill% \begin{minipage}[t]{\halfpagewid} \begin{tabular*}{\halfpagewid}[t]{@{}l@{\extracolsep{\fill}}r@{}} {\tt PYOFSH} subroutine & \pageref{p:PYOFSH} \\ {\tt PYPARS} common block & \pageref{p:PYPARS1}, \pageref{p:PYPARS2} \\ {\tt PYPILE} subroutine & \pageref{p:PYPILE} \\ {\tt PYQQBH} subroutine & \pageref{p:PYQQBH} \\ {\tt PYRAND} subroutine & \pageref{p:PYRAND} \\ {\tt PYREMN} subroutine & \pageref{p:PYREMN} \\ {\tt PYRESD} subroutine & \pageref{p:PYRESD} \\ {\tt PYSCAT} subroutine & \pageref{p:PYSCAT} \\ {\tt PYSIGH} subroutine & \pageref{p:PYSIGH} \\ {\tt PYSPEN} function & \pageref{p:PYSPEN} \\ {\tt PYSPLI} subroutine & \pageref{p:PYSPLI} \\ {\tt PYSSPA} subroutine & \pageref{p:PYSSPA} \\ {\tt PYSTAG} subroutine & \pageref{p:PYSTAG} \\ {\tt PYSTAT} subroutine & \pageref{p:PYSTAT} \\ {\tt PYSTEL} subroutine & \pageref{p:PYSTEL} \\ {\tt PYSTFL} subroutine & \pageref{p:PYSTFL} \\ {\tt PYSTFU} subroutine & \pageref{p:PYSTFU} \\ {\tt PYSTGA} subroutine & \pageref{p:PYSTGA} \\ {\tt PYSTGS} subroutine & \pageref{p:PYSTGS} \\ {\tt PYSTPI} subroutine & \pageref{p:PYSTPI} \\ {\tt PYSTPR} subroutine & \pageref{p:PYSTPR} \\ {\tt PYSUBS} common block & \pageref{p:PYSUBS} \\ {\tt PYTEST} subroutine & \pageref{p:PYTEST} \\ {\tt PYUPEV} subroutine & \pageref{p:PYUPEV} \\ {\tt PYUPIN} subroutine & \pageref{p:PYUPIN} \\ {\tt PYUPPR} common block & \pageref{p:PYUPPR} \\ {\tt PYWAUX} subroutine & \pageref{p:PYWAUX} \\ {\tt PYWIDT} subroutine & \pageref{p:PYWIDT} \\ {\tt PYXTOT} subroutine & \pageref{p:PYXTOT} \\ {\tt Q2UP} in {\tt PYUPPR} & \pageref{p:Q2UP} \\ {\tt RLU} function & \pageref{p:RLU} \\ {\tt RLUGET} subroutine & \pageref{p:RLUGET} \\ {\tt RLUSET} subroutine & \pageref{p:RLUSET} \\ {\tt RRLU} in {\tt LUDATR} & \pageref{p:RRLU} \\ {\tt SIGH} in {\tt PYINT3} & \pageref{p:SIGH} \\ {\tt SIGT} in {\tt PYINT7} & \pageref{p:SIGT} \\ {\tt ULALEM} function & \pageref{p:ULALEM} \\ {\tt ULALPS} function & \pageref{p:ULALPS} \\ {\tt ULANGL} function & \pageref{p:ULANGL} \\ {\tt ULMASS} function & \pageref{p:ULMASS} \\ {\tt V} in {\tt LUJETS} & \pageref{p:V} \\ {\tt VCKM} in {\tt LUDAT2} & \pageref{p:VCKM} \\ {\tt VINT} in {\tt PYINT1} & \pageref{p:VINT} \\ {\tt WIDE} in {\tt PYINT4} & \pageref{p:WIDE} \\ {\tt WIDP} in {\tt PYINT4} & \pageref{p:WIDP} \\ {\tt WIDS} in {\tt PYINT4} & \pageref{p:WIDS} \\ {\tt XSEC} in {\tt PYINT5} & \pageref{p:XSEC} \\ {\tt XSFX} in {\tt PYINT3} & \pageref{p:XSFX} \\ \end{tabular*} \end{minipage} \end{document}