Table of Contents Author Guidelines Submit a Manuscript
Advances in High Energy Physics
Volume 2019, Article ID 3142510, 17 pages
https://doi.org/10.1155/2019/3142510
Review Article

A Short Review on Recent Developments in TMD Factorization and Implementation

Departamento de Física Teórica and IPARCOS, Universidad Complutense de Madrid, Ciudad Universitaria, 28040 Madrid, Spain

Correspondence should be addressed to Ignazio Scimemi; se.mcu@soizangi

Received 20 December 2018; Accepted 7 April 2019; Published 13 May 2019

Guest Editor: Zhongbo Kang

Copyright © 2019 Ignazio Scimemi. This is an open access article distributed under the Creative Commons Attribution License, which permits unrestricted use, distribution, and reproduction in any medium, provided the original work is properly cited. The publication of this article was funded by SCOAP3.

Abstract

In the latest years the theoretical and phenomenological advances in the factorization of several collider processes using the transverse momentum dependent distributions (TMD) have greatly increased. I attempt here a short resume of the newest developments discussing also the most recent perturbative QCD calculations. The work is not strictly directed to experts in the field and it wants to offer an overview of the tools and concepts which are behind the TMD factorization and evolution. I consider both theoretical and phenomenological aspects, some of which have still to be fully explored. It is expected that actual colliders and the Electron Ion Collider (EIC) will provide important information in this respect.

1. Introduction

The knowledge of the structure of hadrons is a leitmotiv for the study of quantum chromodynamics (QCD) for decades. Apart from the notions of quarks and gluons (we call them generically “partons” in the following), the natural question is how the momenta of these particles are distributed inside the hadrons and how the spin of hadrons is generated. Phenomenologically it is possible to access at this problem only in some particular kinematical conditions, as provided, for instance, in experiments like (semi-inclusive) deep inelastic scattering, vector and scalar boson production, hadrons, or jets. I review the basic principle which supports this investigation. Let us consider, to start with, the cross section for di-lepton production in a typical Drell-Yan process where includes all particles which are not directly measured. The cross section for this process can be written formally aswhere is the virtual di-lepton invariant mass, are the parton momenta fraction along a light-cone direction or Bjorken variables, and are the parton distribution functions (PDF). The r.h.s. of (1) assumes several notions which, nowadays, can be found in textbooks. In fact a central hypothesis is a clear energy separation between the di-lepton invariant mass and the scale at which QCD cannot be treated perturbatively any more (we call it the hadronization scale GeV), that is, . Given this, one can factorize the cross section in a perturbatively calculable part and the rest. Formula (1) represents just a first term of an “operator product expansion” of the cross section. The price to pay for this separation is the introduction of a factorization scale which can be used to resume logarithms in combination with renormalization group equations [13]. Another aspect, which is remarkable, is that the nonperturbative part of the cross section can be also expressed as the product of two parton distribution functions. This fact has two main consequences: on the one hand, all the nonperturbative information of the process is included in the PDFs; on the other hand, the partons belonging to different hadrons are completely disentangled. In these conditions so the longitudinal momenta of quarks and gluons can be reconstructed nonperturbatively and this fact has given rise to a large investigation whose review goes beyond the purpose of this writing.

The ideal description of the process in (1) however becomes more involved in the case of more differential cross sections [3234]. So, for instance, one can wonder whether a formula likehas any physical consistency. (I use the notation for 2-dimensional impact parameter, , is the center of mass energy of the process, ) The answer to this question is necessarily more complex than in the case of (1) for the simple fact that a new kinematic scale, , the transverse momentum of the di-lepton pair, has now appeared. In this article I will concentrate on the description of the casewhich is interesting for a number of observables. The restriction to this kinematical regime represents also a limitation of the present approach which should be overcome with further studies.

The study of factorization [25, 27, 29, 30, 35, 36] has lead finally to the conclusion that actually (2) in not completely correct because the cross section for these kind of processes should instead be of the formwith and being the rapidity scales. Formula (4) shows explicitly that the TMD functions contain nonperturbative QCD information different from the usual PDF, while they still allow completing disentangle QCD effects coming from different hadrons. These new nonperturbative QCD inputs can be written in terms of well-defined matrix elements of field operators which can be extracted from experiments or evaluated with appropriate theoretical tools. These objectives require some discussion, which I partially provide in this text.

The scale is the authentic key stone of the TMD factorization. Its origin is different from the usual factorization scale and because of this it is allowed to perform a special resummation for this scale. This leads to the fact that a consistent and efficient implementation of the evolution is crucial for the prediction and extraction of TMDs from data. A possible implementation of the TMD evolution is historically provided by Collins-Soper-Sterman (CSS) [3234]. However a complete discussion of more efficient alternatives has started more recently [2123, 26, 37]. The point is that the rapidity scale evolution has both a perturbative and nonperturbative input, as it is actually provided by (derivatives of) an operator matrix element (the so called soft function). An efficient implementation and scale choice so should separate as much as possible the nonperturbative inputs with different origin inside the cross sections. This target is not completely realized with the CSS implementation, while it can be achieved with the -prescription discussed in the text. This discussion is also relevant for multiple reasons. In fact various orders in perturbation theory are available already for unpolarized and polarized distribution and, in the future, one expects more results in this respect for many polarized distributions. When dealing with several perturbative orders, the convergence of the perturbative series can be seriously undermined by an inappropriate choice of scales, and this is a well-known problem that can affect the theoretical error of any result. A more subtle issue comes from the fact that the evolution corrections can also be of nonperturbative nature. It would be certainly clarifying a scheme in which the nonperturbative effects of the evolution are clearly separated from the intrinsic nonperturbative TMD effects. Such a request results to be important when several extraction of TMD from data are compared and also when a complete nonperturbative evaluation of TMD can be provided.

In the rest of this review I will try to give an idea on how all these problems can be consistently treated, which can be useful also to explore new and more efficient solutions. Several parts of this review use material that can be originally found in [4, 23, 38].

2. Factorization

The factorization of the cross sections into TMD matrix elements has been provided by several authors and it has been object of many discussions [25, 27, 29, 30, 3236]. We briefly review the main ideas here for the case of Drell-Yan. The process is characterized by two initial hadrons which come from opposite collinear directions and produce two leptons in the final state plus unmeasured radiation. We identify collinear (anticollinear) light-cone directions () and , for the momentum of colliding particles. The momentum of collinear particles is with and and . The momenta of collinear particles are characterized by the scaling where is the di-lepton invariant mass and is a small parameter being the hadronization scale. A reversed scaling of momentum is valid for anticollinear particles, say . The soft radiation which entangles collinear and anticollinear particles is homogeneous in momentum distribution (its momentum scales as ) and can be distinguished from the collinear radiation only for a different scaling of the components of the momenta. Given this, it is natural to divide the hadronic phase space in regions as in Figure 1. In this picture, the collinear and soft regions are necessarily separated by rapidity and they all share the same energy .

Figure 1: Diagrams of regions for TMD factorization (original figure in [30]).
2.1. Soft Interactions and Soft Factor

Because the soft radiation is not finally measured, its interactions should be included (and resumed) in the collinear parts, which become sensitive to a rapidity scale which acts in a way similar to the usual factorization scale. It is possible to define the soft radiation through a “soft factor”; that is, by an operator matrix element,where we have used the Wilson line definitions [3941] appropriate for a Drell-Yan process,

The direct calculation of the soft factor is all but trivial and the way the calculation is performed can influence directly the final formal definition of the transverse momentum dependent distribution used by different authors. In fact a simple perturbative calculation shows that in the soft factor there are divergences which cannot be regularized dimensionally (say, they are not explicitly ultraviolet (UV) or infrared (IR)) which occur when the integration momenta are big and aligned on the light-cone directions. The divergences that arise in this configuration of momenta are generically called rapidity divergences and regulated by a rapidity regulator. One can understand the necessity of a specific regulator observing that the light-like Wilson lines are invariant under the coordinate rescaling in their own light-like directions. This invariance leads to an ambiguity in the definition of rapidity divergences. Indeed, the boost of the collinear components of momenta , (with an arbitrary number) leaves the soft function invariant, while in the limit one obtains the rapidity divergent configuration. Therefore the soft function cannot be explicitly calculated without a regularization which breaks its boost invariance. The coordinate space description of rapidity divergences as well as the counting rules for them have been derived in [42, 43]. The nature of the divergences in the soft factor has been studied explicitly in [44] at one loop and in [45] at NNLO, which conclude that, once all contributions are included, the soft factor depends only on ultraviolet and rapidity divergences (and IR divergences are present only in the intermediate steps of the calculations, but not in the final result). Different regulators have also been shown to be more or less efficient within different approaches to the calculations of transverse momentum dependent distributions. For instance, NNLO perturbative calculations for unpolarized distributions, transversity, and pretzelosity have been performed using de -regulator of [4, 15, 19] while for the recent attempts of lattice calculations off-the-light-cone Wilson lines are preferred [4656]. The discussion of the type of regulator involves usually another issue, which is also important for the complete definition of TMDs. While collinear and soft sectors can be distinguished by rapidity, the choice of a rapidity regulator forces a certain overlap of the two regions which should be removed, in order to arrive to a consistent formulation of the factorized cross section. This is called “zero-bin” problem in Soft Collinear Effective Theory (SCET) [57]) and its solution is usually provided in any formulation of the factorization theorem. The amount of the zero-bin overlap is usually fixed by the same soft function in some particular limit although it is generally impossible to define this subtraction in a unique (in the sense of regulator independent) form. Because of this overlap one can find in the literature that the soft function is used in a different way in different formulations of the factorization theorem. The evolution properties of TMDs however are independent of these subtleties and they are the same in all formulations. A possible rapidity renormalization scheme-dependence is traditionally fixed by requiring (for this notation see discussion on Section 2.2).

The factorization theorem to all orders in perturbation theory relies on the peculiar property of soft function of being at most linear in the logarithms generated by the rapidity divergences. Then it comes natural to factorize it in two pieces [30], and in turn this feature allows to define the individual TMDs. Using the -regulator one can write to all orders in perturbation theory, as well as to all orders in the -expansion (the UV divergences are regulated in dimensional regularization )[45].where tildes mark quantities calculated in coordinate space, is an arbitrary and positive real number that transforms as under boosts, and we introduce the convenient notation Despite the fact that the soft function is not measurable per se, its derivative provides the so-called rapidity anomalous dimension,with . Because of its definition the rapidity anomalous dimension has both a perturbative (finite; calculable) part and a nonperturbative part. This fact should be always taken into account despite the fact that many experimental data are actually marginally sensitive to the nonperturbative nature of the rapidity anomalous dimension. A nonperturbative estimation of the evolution kernel with lattice has been recently proposed in [58] and I expect a deep discussion on this issue in the future. A renormalon based calculation has also provided some approximate value for this nonperturbative contribution [59].

2.2. TMD Operators

Another fundamental ingredient in the formulation of the factorization theorem is represented by the definition of the TMD operators that are involved. We use here the notation of [4]. The TMDs which appear in a Drell-Yan process can be rewritten starting from the bare operators (here I consider only the quark case, for simplicity)where , and are light-cone vectors (), and is some Dirac matrix; the repeated color indices () are summed up. The representations of the color SU(3) generators inside the Wilson lines are the same as the representation of the corresponding partons. The collinear Wilson lines are defined in the same way as in in (6). The collinear and soft Wilson line should be distinguished because the gluons which define them have a different scaling in effective field theories and also because they should be regularized differently with respect to rapidity divergences (see the definition of -regulator in [45] for soft and collinear matrix elements). The hadronic matrix elements,define the bare or unsubtracted TMDs. These bare operators do not include for the moment any soft radiation and they are just collinear object (one can refer to them as “beam functions”) and because of boost invariance they can be calculated in principle in any frame. A renormalization issue however appears because of rapidity divergences and overlap with the soft radiation (this problem is usually referred to as zero-bin problem in effective field theory [57]). The soft interactions can be incorporated in the definition of the TMD through an appropriate “rapidity renormalization factor”. The final form of the rapidity renormalization factor ( in the following) is dictated by the factorization theorem. The renormalized operators and the TMD are defined, respectively, asand is the UV renormalization constant for TMD operators and is the rapidity renormalization factor. Both these factors are the same for particle and antiparticle; however they are different for quarks and gluons. These factors also occur in the same way in parton distribution functions and fragmentation functions. The scales and are the scales of UV and rapidity subtractions, respectively. The way these factors are ordered corresponds to first removing rapidity divergences and then the rest of UV divergences from the bare matrix elements as in [4]. It is possible to proceed also in a different way (for instance, in [29, 60, 61] they cancel the rapidity divergences from the beam functions and soft factors independently); however finally one achieves an equivalent resummation of rapidity logarithms. In [5, 27, 62] for TMDPDFs the soft function is hidden in the product of two TMDs.

Some comments finally are necessary for the zero-bin overlap problem. In principle an overlap factor affects the rapidity renormalization factor aswhere is the soft function and is the zero-bin contribution [25, 30, 36, 57, 63] and both are different in the quark and gluon cases. The zero-bin part assumes a particular form depending on the regulator for rapidity divergences. For instance, the modified -regularization [45] has been constructed such that the zero-bin subtraction is literally equal to the soft function: . The definition is nontrivial because it implies a different regularized form for collinear Wilson lines and for soft Wilson lines . In the modified -regularization, the expression for the rapidity renormalization factor isand this relation has been tested at NNLO in [4, 24, 45]. In the formulation of TMDs by Collins in [25] the rapidity divergences are handled by tilting the Wilson lines off-the-light-cone. Then the contribution of the overlapping regions and soft factors can be recombined into individual TMDs by the proper combination of different soft functions with a partially removed regulator. This combination gives the factor ,The rest of logical steps remain the same as with the -regulator. Notice that, due to the process independence of the soft function [25, 30, 36, 63, 64], the factor is also process independent.

An important aspect of factorization is finally represented by the cancellation of unphysical modes, the Glauber gluons. A check of this cancellation has been provided in [25, 6567] and I do not review it here.

3. Matching at Large (or Small-)

The practical implementation of the TMD for data analysis benefits from asymptotic limits of the distribution. These limits allow defining the TMDs at different scale and constraining the nonperturbative behavior of the TMDs. Commonly one starts with the large transverse momentum limit of the TMD. In this case one can refactorize the TMDs in terms of Wilson coefficient and collinear parton distribution functions (PDF), following the usual rules for operator product expansion (OPE). At operator level one findsand the symbol is the Mellin convolution in variable or , and enumerate the flavors of partons. The running on the scales , , and is independent of the regularization scheme and it is dictated by the renormalization group equations which we discuss later. In the case of initial states (17) reduces towhere are the integrated collinear distributions, that is, the parton distribution functions (PDF) which depend only on the Bjorken variables ( for PDFs) and the renormalization scale . All dependence on the transverse coordinate and rapidity scale is contained in the matching coefficient and can be calculated perturbatively. I report the definition of the PDFs for completeness

The calculation of the Wilson coefficients in (17) uses the standard methods of the operator product expansion which profit of the fact that the coefficients do not depend on the infrared limit of the matrix elements. The current status of these calculations for quark distributions is resumed in Table 1. Less information is generally available in the case of gluon TMDs. Basically the matching coefficients for unpolarized gluons are known at NNLO [6] and linearly polarized gluons at NLO [15]. In general the TMDs which match onto collinear twist-3 functions are much less known, which reflects the difficulty of the computations. It would be very useful to have a better knowledge of all these less known functions at higher perturbative order before the advent of Electron Ion Collider (EIC). In the rest of this Section 1 focuses on unpolarized quark distributions which offer also an important understanding on the power of the TMD factorization. The necessity of a complete NLO estimation of all TMDs is both theoretical and phenomenological. Actually a difficulty of the TMD extraction from data is due to the fact that it is a nontrivial function of two variables (Bjorken and transverse momentum) so that a complete mapping on a plane is necessary. This target is achievable thanks to the factorization of the cross section and the consequent extraction of the TMD evolution part, which is process independent. A second important information comes from the asymptotic limit of the TMD for large transverse momentum, which is perturbatively calculable. The simple LO expressions for the TMD in general do not provide much information (they are just constants), so that in order to achieve a wise modeling a NLO calculation is always necessary. The higher order calculations allow also testing the stability with respect to the scales that match the TMD perturbative and nonperturbative parts. For the unpolarized case a study in this sense can be found in [22] both for high energy and low-energy data. Using a LO calculations one cannot even quantify this error. Finally, another lesson that comes from the analysis of the unpolarized case is that a good portion of the TMD is tractable starting from their asymptotic expansion for large transverse momenta. In any case even a 10% average precision of the SIDIS cross section at EIC will need a NLO theoretical input.

Table 1: Summary of available perturbative calculations of quark TMD distributions and their leading matching at small-b.

4. Evolution

The evolution of the TMDs in the factorization, , scale is derived from their defining operators and from (15),in an usual way. Equation (20) is a standard renormalization group equation (which comes from the renormalization of the ultraviolet divergences), the function is called the TMD anomalous dimension, and it contains both single and double logarithms. The same (13) can be used to write the running with respect to the rapidity scale, . Because the rapidity scale evolution comes from soft interactions and more specifically from the soft factor (see discussion in [45, 61] and, e.g., [24, 42, 43]) which is the same for initial and final states, the rapidity scale evolution is the same for TMD parton distribution functions and TMD fragmentation functions, and it is also spin-independent (so it is the same also for TMDs at higher twist),The function is called the rapidity anomalous dimension and actually one has . The anomalous dimensions for these pair of evolution have been addressed with several names in the literature as it is shown in Table 2.

Table 2: Notation for TMD anomalous dimensions used in the literature.

Quark and gluon rapidity anomalous dimensions are related up to three loops by the Casimir scaling (see [45]),

The consistency of the differential equations ((20)-(21)) implies that the cross-derivatives of the anomalous dimension are equal to each other ([45, 61]),From (22) one finds that the anomalous dimension is where we introduce the notationThe large- expansion of the TMD introduces also another evolution scale, which is needed for in the matching of Wilson coefficients with the collinear operators. In the case of the unpolarized TMDs this is provided by the DGLAP (DGLAP is an acronym for Dokshitzer, Gribov, Lipatov, Altarelli, Parisi [13]) equations where are the DGLAP kernels for the PDF. Similar equations hold for unpolarized TMD fragmentation functions. It is useful to recall also the running of the matching coefficient with respect to the rapidity scale (we set ) The solutions of these differential equations are This defines the reduced matching coefficients whose renormalization group evolution equations are with the kernel The perturbative expansion of all these functions provides consistency requirements for the logarithmic terms at a given order. Using the notation for the -th perturbative order,one finds that the knowledge of the coefficient at order provides all the terms with at order in this series. So finally any higher order calculation provides new information on terms . A resume of the present status of available calculations is provided in Table 1.

5. Implementation of TMD Formalism and TMD Extraction from Data

As an example of application of the TMD formalism I review the study of unpolarized TMD parton distribution functions in Drell-Yan and Z-boson production following [23].

Namely, I consider the process , where is the electroweak neutral gauge boson, or . The incoming hadrons have momenta and with . The gauge boson decays to the lepton pair with momenta and . The momentum of the gauge boson or equivalently the invariant mass of lepton pair is . The differential cross section for the Drell-Yan process can be written in the form [68, 69]where is the flux factor; is the (Feynman) propagator for the gauge boson . The hadron and lepton tensors are, respectively,where is the electroweak current. Within the TMD factorization, the unpolarized hadron tensor (see, e.g., [70]) iswhere is the transverse part of the metric tensor and the summation runs over the active quark flavors. The variable is the hard factorization scale. The variables are the scales of soft-gluons factorization, and they fulfill the relation . In the following, we consider the symmetric point . The factors are the electro-weak charges and they are given explicitly in [23]. The factor is the matching coefficient of the QCD neutral current to the same current expressed in terms of collinear quark fields. The explicit expressions for can be found in [7173].

In (34) I have not included power corrections to the TMD factorization (to be distinguished from the power corrections to the TMD operator product expansion). It is difficult to establish the amount of these corrections but a phenomenological study in [23] and a more formal study in the large- limit (that is, the limit of large number of colors) in [74] have found a reasonable upper value . A study which takes into account the structure of operators in the type of corrections has been started in [75]. In general the power corrections should be included when the di-lepton invariant mass is of order a few GeV (this is the case, for instance, of HERMES experiment and, perhaps to a possibly less extent, COMPASS) or when the experimental precision is extreme (as it possibly happens with ATLAS experiment). This is issue is important phenomenologically and involves the study of cross sections with the inclusion of factorization breaking contributions. Some recent suggestion have appeared in [76, 77] which have still to be tested phenomenologically. One should remark however that the implementation of these factorization breaking corrections strongly depends on the fact that the factorized part of the cross section is correctly realized and phenomenologically tested. More studies on this issue are necessary in the future.

Evaluating the lepton tensor and combining together all factors one obtains the cross section for the unpolarized Drell-Yan process at leading order of TMD factorization, in the form [25, 27, 34, 36, 78, 79]where is the rapidity of the produced gauge boson. The factor is a part of the lepton tensor and contains information on the fiducial cuts. This factor provides important information on the actual measured leptons and should be always included when the relative experimental information is provided.

The evolution of the TMDs play a special role in (35) and we collect of evolution equations here:and on the right hand side of these equation we have omitted the reference to flavor for simplicity. The main features of these anomalous dimensions that have been discussed in previous sections are the following: The TMD anomalous dimension contains both single and double logarithms and the anomalous dimension refers to the finite part of the renormalization of the vector form factor; see Table 2. Equation (36) cannot fix the logarithmic part of entirely, but only order by order in perturbation theory, because the parameter is also responsible for the running of the coupling constant. The rapidity anomalous dimension is a nonperturbative function (see the discussion about the renormalon in the perturbative series of this function in [32, 59, 80]), although it can be perturbatively expanded for small .

The double-evolution equation of the TMDs can be formulated as in [23] using a two-dimensional vector field notation, that i reproduces here. The procedure consists in introducing a convenient two-dimensional variable which treats scales and equally,where the dimension of the scale parameters is explicitly indicated and the bold font means the two-dimensional vectors. Then one defines the standard vector differential operations in the plane , namely, the gradient and the curlThe TMD anomalous dimensions can be all included in a vector evolution field , Here and in the following, we use the vectors as the argument of the anomalous dimensions for brevity, keeping in mind that , , etc. In other words, the anomalous dimensions are to be evaluated on the corresponding values of and defined by value of in (37). The TMD evolution equations (36) and the evolution factor in this notation have the formUsing this formalism, (36) are equivalent to the statement that the evolution flow is irrotational,and is the evolution scalar potential for TMD. According to the gradient theorem any line integral of the field is path-independent and equals the difference of values of potential at end-points. Therefore, the solution for the factor in (40) isand are the first and second components of the vector in (37), and the last term is an arbitrary -dependent function.

The evolution field presented in the previous section is conservative only when the full perturbative expansion of the evolution equations is known. In practice only a few terms of the evolution are calculated, so that it is important to understand in which sense the evolution field remains conservative. Using the Helmholtz decomposition, the evolution field is split into two parts The field is irrotational, the field is divergence-free, and they are orthogonal to each otherwith the notation . Then, one can write the irrotational field , which is conservative, as the gradient of a scalar potentialand the divergence-free part as the vector curl (see (38)) of another scalar potential The curl of the evolution field can be calculated using the definitions (36),The function can be calculated order by order in perturbation theory. For instance, at order one findsis the -function with first terms removed. For instance, we haveIn these expressions the -function is not expanded because in applications one can find a different perturbative order with respect to the rest of the anomalous dimensions.

The immediate consequence of the fact that the evolution field is no more conservative is that the evolution factor is dependent on the path chosen to join the initial and final points and this fact introduces a theoretical error which can be dominant in certain implementation of the evolution kernels. The difference between two solutions evaluated on different paths iswhere is the closed path built from paths and and is the area surrounded by these paths. Using the independence of on the variable , (51) becomes where is the -component of the path at the scale . This equation shows that the difference between paths becomes bigger with largely separated rapidity scales .

The path independence of the evolution is crucial for the implementation of the perturbative formalism, as its absence can derive into uninterpretable extractions of TMDs or big theoretical errors. The path independence can be achieved observing thatshould hold order by order in perturbation theory. Once this is realized it is possible to define null-evolution lines in the plane, which coincide with equipotential lines, and the evolution of TMD takes place only between two different lines. I resume here two possible solutions to this problem, following [23].

In the literature one can find a typical way to implement the evolution that one can call the improved scenario which includes the Collins-Soper-Sterman formalism [21, 22, 25, 26, 29, 81, 82]. In this scenario one chooses a scale such thatIn this way one obtainsand the TMD evolution factor depends on The situation in this scenario can be visualized in Figure 2. Any choice of corresponds to a different scheme as is the point where evolution flips from path 1 and path 2 in Figure 2. The differences that can appear in the extraction of TMDs which depend on the choice of can be numerically large, so that the selection of this scale can cause also some problems when a sufficient precision is required.

Figure 2: Paths for the improved solution which depend on the choice of the reference scale .

The presence of the intermediate scale is not unavoidable in the implementation of the TMD evolution. In fact the integrability condition (53) can be restored by changing the anomalous dimension to a modified value such thatand the corresponding solution for the evolution factor readsThese expressions should be completed with the resummation of by means of renormalization group equation (55) as it is not implicitly included in this scenario.

5.1. Prescription and Optimal TMDs

Because we have a double scale running of the TMD one can find lines of null evolution in the plane. These lines are by default the equipotential lines, that we call . In formulas this can be written aswhen the scales and belong to the same null-evolution curve. Any point of the line in the plane does not change the value of the TMDs. When two TMDs do not belong to the same equipotential curve one findswhere is defined such that and . In order to minimize the evolution effect and so to have a more stable prediction/extraction of TMDs the initial and final evolution curves should be selected with care. Once this is settled, it remains to find an appropriate set of initial and final line which is sufficiently stable for all relevant processes. The final point of the rapidity evolution, , is as usual dictated by the hard subprocess. Concerning the starting line it is convenient to take into account also the matching of the TMD on the respective integrated distributions, which formally iswhere is PDF or FF and is the Wilson coefficient function. The coefficient function includes the dependence on within the logarithms and . The traditional choice, see, e.g., [25, 26, 83], leaves uncanceled logarithmic factors in the coefficient function which explode at small . The damage caused by this choice is partially cured inserting more prescriptions like the prescription [25], which however bias the modeling of the nonperturbative part of the TMD. In fact, among the others, this prescription correlated the nonperturbative part of the evolution factor with the intrinsic nonperturbative part of the TMD. Although this prescription may work for some initial studies it results to have serious drawbacks for more precise analysis.

The -prescription suggested in [22, 23] provides an attempt to improve the stability of the perturbative series and keep separate the truly nonperturbative TMD distribution from the nonperturbative part of the evolution kernel. The advantage of this separation is that one can always use all known perturbative information for the evolution factor, even if the knowledge of the collinear matching is not at the same perturbative order. This reduces drastically the perturbative uncertainty in the extractions of TMDs from data and also facilitates the understanding of direct calculations of the TMD through nonperturbative QCD methods like lattice.

I provide here some description of the -prescription following [22, 23]. A TMD distribution in the -prescription readswhere is defined such that , that is, the value of is such that the initial scales of the TMD distribution belong to the null-evolution curve . At this stage let me rewrite (61) specifying the scales,where is an intrinsic scale for the expansion of the TMD in terms of Wilson coefficients and PDFs and it is a free parameter. In general the values of are restricted by the values of ,except

The last choice give us much more freedom to model the nonperturbative part of the TMD and the definition of the initial scale is unique and nonperturbatively defined. The choice of as the initial point is so optimal and consistent with the reexpression of TMDs using PDFs. This scheme fixes the optimal-TMD-distribution; that is, it fixes the initial special null-evolution curve. As a summery any initial point in the saddle curve is defined nonperturbatively and it is unique and performing this choice it is consistent to write optimal TMD simply as .

A plot for the -factor is given in Figure 3. At large- the shape of the rapidity anomalous dimension is nonperturbative as renormalon studies confirm [59, 80] (see also [84]). So, at large- the expression for should be extracted from data fitting, while at small- it should match the perturbative expression. We recall that the -prescription has, among its benefit, the one of separating the modeling of the nonperturbative part of the evolution factor from the rest of nonperturbative parts of the TMD. This implies that it always recommendable to use the highest nonperturbative order in the evolution factor, even if the matching coefficients of the TMD with the collinear functions are known to a lesser precision perturbatively.

Figure 3: evolution factor using the optimal TMD prescriptions when the high scale is fixed at the Z-boson mass (right side) and at BELLE center of mass energy 10.52 GeV (left side). In this figure one has chosen GeV and with GeV.

The nonperturbative part of the evolution kernel can be modeled in different ways. For instance, one can introduce a simple ansatz like a the modificationwhere and is a parameter, such that as suggested a long ago in [33], as part of the prescription [25]. Let us stress that the choice of a can be admissible separately for the evolution factor and that (66) does not imply -prescription for the whole TMD distribution. With the choice the saddle point is always in the observable region, which allows determining the optimal TMD. In this case the evolution factor readsand the resumed expression for the TMD anomalous dimension as in [21] is understood in the last line. In (68), the scale is -dependent and defined by the equation

The expression for the cross section with the optimal TMD definition is particularly compact and reads

The derivation of the saddle point using formula (69) is in practice done numerically, so that an efficient method to extract it or to approximate this point should be discussed as in [23]. A technical discussion of this issue is beyond the point of this paper.

Let me conclude this section discussing a comparison of the optimal TMD construction with a more traditional implementation on data following the recent fits in [22, 85]. The absence of an intermediate scale removes one (scheme dependent) source of error and at the same time it allows the path independence of the final result. In this way it is possible to directly compare from different extractions and models. In the definition of the optimal TMD the low-energy normalization is defined “nonperturbatively” and uniquely by (69) which implies that the perturbative order of the evolution is completely unrelated to the perturbative order of matching of the TMD on the respective collinear functions. Because the evolution factor is known often at higher orders with respect to the Wilson coefficient matching factors, it is always possible to fully use all the available perturbative information. The theoretical uncertainty of TMDs is estimated with the variation of and . The fact that the number of varied scales is different from more standard analysis does not necessarily imply a reduction of theoretical errors. The error in fact reshuffles in and , but the description is now more coherent. One can appreciate this effect in Figure 4 taken from [22]. In this figure one compares for the ATLAS experiment a standard method to test the dependence on the scales and thus the stability of the perturbation theory prediction, multiplying each scale by a parameter [22, 37, 86, 87], and varying the parameters nearby their central value. For example, in the notation of [22], one changes scales asand checks the variations of . The variation of all these four parameters is consistent with a nonoptimal definition of TMDs, while in the optimal case only the variation of and is necessary.

Figure 4: Comparison of error bands obtained by the scale-variations for cross sections at NNLO in traditional (upper figure) and optimal (lower figure) TMD implementation. Here, the kinematics bin-integration, etc., is for the Z-boson production measure at ATLAS at 8 TeV [31].

6. Conclusions

The formulation of factorization theorems in terms of TMDs is a first fundamental step for the study of the structure of hadrons and the origin of spin. The use of the effective field theory appears essential to correctly order the QCD contributions. Properties of TMDs like evolution and their asymptotic limit at large values of transverse momentum can be systematically calculated starting from the definition of correct operators and the evaluation of the interesting matrix elements. A key point for the renormalization of TMDs is represented by the so-called soft matrix element which is common in the definition of all spin dependent leading twist TMD.

Still, all this is just a starting point for the study of TMDs. In fact a correct implementation of evolution requires a control of all renormalization scales that appear in the factorization theorem. I have described here some of these possibility putting the accent on some recent interesting developments which, at least theoretically, allow a better control of the resumed QCD series. The understanding of factorization allows also precisely defining the range of ideal experimental conditions where this formalism can be applied. A full analysis of present data using all the theoretical information collected so far is still missing and it will certainly be an object of research in the forthcoming years. The formalism described in this work is the one developed for unpolarized distributions. However the evolution factors are universal; that is, they are the same for polarized and unpolarized leading twist TMDs and they are valid in Drell-Yan, SIDIS experiments, and colliders, where the factorization theorem applies. All this formalism is expected to be tested on data in the near future. Nevertheless a lot of perturbative and nonperturbative information is still missing. Giving a look at Table 1 one can see that for many TMD one has only a lowest order perturbative calculation which should be improved in order to have a reliable description of data. While the information on the nonperturbative structure of TMD is still poor and still driven by phenomenological models, it is important to implement the TMD formalism in such a way that perturbative and nonperturbative effects are well separated. And among the nonperturbative effects, one should be able to distinguish the ones of the evolution kernel from the rest. In the text I have discussed a possible solution to this problem. Some prominent research lines which possibly will deserve more attention in the future include the cases where hadrons are measured inside the jets, see, for instance, [8890] or outside a jet (say, hadron-jet interactions) [9193] and lattice.

Conflicts of Interest

The authors declare that they have no conflicts of interest.

Acknowledgments

I would like to thank Alexey Vladimirov and Daniel Gutierrez Reyes for discussing this paper. Ignazio Scimemi is supported by the Spanish MECD grant FPA2016-75654-C2-2-P.

References

  1. V. N. Gribov and L. N. Lipatov, “Deep inelastic e p scattering in perturbation theory,” Soviet Journal of Nuclear Physics, vol. 15, pp. 438–450, 1972, [Yad. Fiz.15,781(1972)]. View at Google Scholar
  2. Y. L. Dokshitzer and Sov. Phys. JETP, “Calculation of structure functions of deep-inelastic scattering and e+e- annihilation by perturbation theory in quantum chromodynamics,” Journal of Experimental and Theoretical Physics, vol. 46, p. 641, 1977. View at Google Scholar
  3. G. Altarelli and G. Parisi, “Asymptotic freedom in parton language,” Nuclear Physics B, vol. 126, pp. 298–318, 1977. View at Publisher · View at Google Scholar
  4. M. G. Echevarria, I. Scimemi, and A. Vladimirov, “Unpolarized Transverse Momentum Dependent Parton Distribution and Fragmentation Functions at next-to-next-to-leading order,” Journal of High Energy Physics, vol. 2016, p. 4, 2016. View at Publisher · View at Google Scholar
  5. T. Gehrmann, T. Luebbert, and L. L. Yang, “Calculation of the transverse parton distribution functions at next-to-next-to-leading order,” Journal of High Energy Physics, vol. 2014, no. 6, p. 155, 2014. View at Publisher · View at Google Scholar
  6. I. Scimemi and A. Vladimirov, “Matching of transverse momentum dependent distributions at twist-3,” High Energy Physics - Phenomenology, vol. 78, p. 802, 2018. View at Google Scholar
  7. D. Boer, P. J. Mulders, and F. Pijlman, “Universality of T-odd effects in single spin and azimuthal asymmetries,” High Energy Physics - Phenomenology, vol. 667, no. 1-2, Article ID 0303034, pp. 201–241, 2003. View at Publisher · View at Google Scholar
  8. X. Ji, J. W. Qiu, W. Vogelsang, and F. Yuan, “Unified picture for single transverse-spin asymmetries in hard-scattering processes,” Physical Review Letters, vol. 97, Article ID 082002, 2006. View at Publisher · View at Google Scholar
  9. X. Ji, J. W. Qiu, W. Vogelsang, and F. Yuan, “Single transverse-spin asymmetry in Drell-Yan production at large and moderate transverse momentum,” Physical Review D: Particles, Fields, Gravitation and Cosmology, vol. 73, Article ID 094017, 2006. View at Publisher · View at Google Scholar
  10. Y. Koike, W. Vogelsang, and F. Yuan, “On the relation between mechanisms for single-transverse-spin asymmetries,” High Energy Physics - Phenomenology, vol. 659, no. 5, pp. 878–884, 2008. View at Google Scholar
  11. Z. B. Kang, B. W. Xiao, and F. Yuan, “QCD resummation for single spin asymmetries,” Physical Review Letters, vol. 107, Article ID 152002, 2011. View at Publisher · View at Google Scholar
  12. P. Sun and F. Yuan, “Energy evolution for the Sivers asymmetries in hard processes,” Physical Review D: Particles, Fields, Gravitation and Cosmology, vol. 88, Article ID 114012, 2013. View at Publisher · View at Google Scholar
  13. L.-Y. Dai, Z.-B. Kang, A. Prokudin, and I. Vitev, “Next-to-leading order transverse momentum-weighted Sivers asymmetry in semi-inclusive deep inelastic scattering: the role of the three-gluon correlator,” Physical Review D: Covering Particles, Fields, Gravitation, and Cosmology, vol. 92, Article ID 114024, 2015. View at Publisher · View at Google Scholar
  14. I. Scimemi, A. Tarasov, and A. Vladimirov, “Collinear matching for Sivers function at next-to-leading order,” High Energy Physics - Phenomenology, 2019. View at Google Scholar
  15. D. Gutierrez-Reyes, I. Scimemi, and A. A. Vladimirov, “Twist-2 matching of transverse momentum dependent distributions,” Physics Letters B, vol. 769, pp. 84–89, 2017. View at Google Scholar
  16. A. Bacchetta and A. Prokudin, “Evolution of the helicity and transversity Transverse-Momentum-Dependent parton distributions,” Nuclear Physics B, vol. 875, p. 536, 2013. View at Publisher · View at Google Scholar
  17. M. G. A. Bu_ng, M. Diehl, T. Kasemets, and JHEP., “Transverse momentum in double parton scattering: factorisation, evolution and matching,” Journal of High Energy Physics, vol. 2018, p. 044, 2018. View at Google Scholar
  18. K. Kanazawa, Y. Koike, A. Metz, D. Pitonyak, and M. Schlegel, “Operator constraints for twist-3 functions and lorentz invariance properties of twist-3 observables,” Physical Review D: Covering Particles, Fields, Gravitation, And Cosmology, vol. 93, Article ID 054024, 2016. View at Google Scholar
  19. D. Gutierrez-Reyes, I. Scimemi, and A. Vladimirov, “Transverse momentum dependent transversely polarized distributions at next-to-next-to-leading-order,” Journal of High Energy Physics, vol. 2018, no. 172, 2018. View at Publisher · View at Google Scholar
  20. X. P. Chai, K. B. Chen, and J. P. Ma, “A note on Pretzelosity TMD parton distribution,” Physics Letters B, vol. 789, pp. 360–365, 2018. View at Google Scholar
  21. M. G. Echevarría, A. Idilbi, A. Schäfer, and I. Scimemi, “Model-independent evolution of transverse momentum dependent distribution functions (TMDs) at NNLL,” The European Physical Journal C, vol. 73, article 2636, 2013. View at Publisher · View at Google Scholar
  22. I. Scimemi, A. Vladimirov, and Eur., “Analysis of vector boson production within TMD factorization,” The European Physical Journal C, vol. 78, p. 89, 2018. View at Google Scholar
  23. I. Scimemi and A. Vladimirov, “Systematic analysis of double-scale evolution,” Journal of High Energy Physics, vol. 2018, p. 3, 2018. View at Google Scholar
  24. M. G. Echevarria, I. Scimemi, and A. Vladimirov, “Erratum: transverse momentum dependent fragmentation function at next-to–next-to–leading order,” Physical Review D: Particles, Fields, Gravitation and Cosmology, vol. 94, no. 9, 2016. View at Publisher · View at Google Scholar
  25. J. Collins, Foundations of Perturbative QCD, Cambridge University Press, 2013. View at Publisher · View at Google Scholar
  26. S. M. Aybat and T. C. Rogers, “Transverse momentum dependent parton distribution and fragmentation functions with QCD evolution,” Physical Review D: Particles, Fields, Gravitation and Cosmology, vol. 83, no. 11, Article ID 114042, 2011. View at Publisher · View at Google Scholar
  27. T. Becher and M. Neubert, “Drell-Yan production at small qT, transverse parton distributions and the collinear anomaly,” The European Physical Journal C, vol. 71, p. 1665, 2011. View at Google Scholar
  28. T. Becher, M. Neubert, and D. Wilhelm, “Electroweak gauge-boson production at small qT: Infrared safety from the collinear anomaly,” Journal of High Energy Physics, vol. 2012, p. 124, 2012. View at Google Scholar
  29. J.-Y. Chiu, A. Jain, D. Neill, and I. Z. Rothstein, “A formalism for the systematic treatment of rapidity logarithms in Quantum Field Theory,” Journal of High Energy Physics, vol. 2012, p. 84, 2012. View at Publisher · View at Google Scholar
  30. M. G. Echevarría, A. Idilbi, A. Schäfer, and I. Scimemi, “Soft and collinear factorization and transverse momentum dependent parton distribution functions,” The European Physical Journal C, vol. 726, pp. 795–801, 2013. View at Publisher · View at Google Scholar
  31. G. Aad et al., “Measurement of the transverse momentum and distributions of Drell-Yan lepton pairs in proton-proton collisions at √s=8 TeV with the ATLAS detector,” The European Physical Journal C, vol. 76, p. 291, 2016. View at Google Scholar
  32. G. Parisi and R. Petronzio, “Small transverse momentum distributions in hard processes,” Nuclear Physics B, vol. 154, no. 3, pp. 427–440, 1979. View at Publisher · View at Google Scholar
  33. J. C. Collins and D. E. Soper, “Back-to-back jets: Fourier transform from b to kT,” Nuclear Physics B, vol. 197, p. 446, 1982. View at Publisher · View at Google Scholar
  34. J. C. Collins, D. E. Soper, G. F. Sterman, and B. Nucl. Phys, “Transverse momentum distribution in Drell-Yan pair and W and Z boson production,” Nuclear Physics B, vol. 250, pp. 199–244, 1985. View at Google Scholar
  35. X. D. Ji, J. Ma, and F. Yuan, “QCD factorization for semi-inclusive deep-inelastic scattering at low transverse momentum,” Physical Review D: Particles, Fields, Gravitation and Cosmology, vol. 71, no. 3, Article ID 034005, 2005. View at Google Scholar
  36. M. G. Echevarria, A. Idilbi, and I. Scimemi, “Factorization theorem for Drell-Yan at low qT and transverse-momentum distributions on-the-light-cone,” Journal of High Energy Physics, vol. 2012, no. 2, 2012. View at Publisher · View at Google Scholar
  37. M. F. Ghajari, M. V. Golpayegani, M. Bargrizan, G. Ansari, and S. Shayeghi, “Non-perturbative QCD effects in qT spectra of Drell-Yan and Z-boson production,” Journal of High Energy Physics, vol. 11, p. 98, 2014. View at Publisher · View at Google Scholar
  38. A. Bacchetta, “Electron-ion collider: the next QCD frontier,” The European Physical Journal A, vol. 2016, p. 163, 2016. View at Google Scholar
  39. A. Idilbi and I. Scimemi, “Singular and regular gauges in soft–collinear effective theory: The introduction of the new Wilson line T,” Physics Letters B, vol. 695, p. 463, 2011. View at Publisher · View at Google Scholar
  40. A. Idilbi and I. Scimemi, “The T‐Wilson Line,” in Proceedings of the Publisher Logo Conference, vol. 1343, p. 320, 2011.
  41. M. García-Echevarría, A. Idilbi, and I. Scimemi, “Soft-collinear effective theory, light-cone gauge, and the T-Wilson lines,” Physical review D, vol. 84, no. 1, Article ID 011502, 2011. View at Publisher · View at Google Scholar
  42. A. Vladimirov and JHEP., “Soft factors for double parton scattering at NNLO,” Journal of High Energy Physics, vol. 2016, p. 38, 2016. View at Publisher · View at Google Scholar
  43. A. Vladimirov, “Structure of rapidity divergences in soft factors,” High Energy Physics - Phenomenology, vol. 4, no. 045, 2018. View at Google Scholar
  44. M. G. Echevarria, A. Idilbi, and I. Scimemi, “On rapidity divergences in the soft and collinear limits of QCD,” International Journal of Modern Physics: Conference Series, vol. 25, Article ID 1460005, 2014. View at Google Scholar
  45. M. G. Echevarria, I. Scimemi, and A. Vladimirov, “Universal transverse momentum dependent soft function at NNLO,” Physical Review D: Particles, Fields, Gravitation and Cosmology, vol. 93, Article ID 054004, 2016. View at Publisher · View at Google Scholar
  46. P. Hagler, B. U. Musch, J. W. Negele, and A. Schafer, “Intrinsic quark transverse momentum in the nucleon from lattice QCD,” High Energy Physics - Lattice, vol. 88, no. 6, Article ID 61001, 2009. View at Publisher · View at Google Scholar
  47. B. U. Musch, P. Hägler, J. W. Negele, and A. Schäfer, “Exploring quark transverse momentum distributions with lattice QCD,” Physical Review D: Particles, Fields, Gravitation and Cosmology, vol. 83, no. 9, Article ID 094507, 2011. View at Publisher · View at Google Scholar
  48. B. U. Musch, P. Hägler, M. Engelhardt, J. W. Negele, and A. Schäfer, “Sivers and Boer-Mulders observables from lattice QCD,” Physical Review D: Particles, Fields, Gravitation and Cosmology, vol. 85, Article ID 094510, 2012. View at Publisher · View at Google Scholar · View at Scopus
  49. X. Ji, “Parton Physics on a Euclidean Lattice,” Physical Review Letters, vol. 110, Article ID 262002, 2013. View at Google Scholar
  50. X. Ji, “Parton physics from large-momentum effective field theory,” Science China Physics, Mechanics & Astronomy, vol. 57, p. 1407, 2014. View at Google Scholar
  51. M. Engelhardt, P. Haegler, B. Musch, J. Negele, and A. Schfer, “Lattice QCD study of the Boer-Mulders effect in a pion,” Physical Review D, vol. 93, Article ID 054501, 2016. View at Google Scholar
  52. B. Yoon, T. Bhattacharya, M. Engelhardt et al., “Lattice QCD calculations of nucleon transverse momentum-dependent parton distributions using clover and domain wall fermions,” in Proceedings of the 33rd International Symposium on Lattice Field Theory (LATTICE2015), pp. 14–18, SISSA, Kobe, Japan, 2015. View at Publisher · View at Google Scholar
  53. B. Yoon, M. Engelhardt, R. Gupta et al., “Lattice QCD calculations of nucleon transverse momentum-dependent parton distributions using clover and domain wall fermions,” High Energy Physics - Lattice, vol. 96, Article ID 094508, 2017. View at Google Scholar
  54. A. V. Radyushkin, “Quasi-parton distribution functions, momentum distributions, and pseudo-parton distribution functions,” Physical Review D: Particles, Fields, Gravitation and Cosmology, vol. 96, Article ID 034025, 2017. View at Publisher · View at Google Scholar
  55. K. Orginos, A. Radyushkin, J. Karpie, and S. Zafeiropoulos, “Lattice QCD exploration of parton pseudo-distribution functions,” Physical Review D: Particles, Fields, Gravitation and Cosmology, vol. 96, no. 9, Article ID 094503, 2017. View at Publisher · View at Google Scholar
  56. X. Ji, L.-C. Jin, F. Yuan, J.-H. Zhang, and Y. Zhao, “Transverse momentum dependent quasi-parton-distributions,” High Energy Physics - Phenomenology, 2018. View at Google Scholar
  57. A. V. Manohar and I. W. Stewart, “The zero-bin and mode factorization in quantum field theory,” High Energy Physics - Phenomenology, vol. 76, Article ID 074002, 2007. View at Publisher · View at Google Scholar
  58. M. A. Ebert, I. W. Stewart, and Y. Zhao, “Determining the nonperturbative Collins-Soper kernel from lattice QCD,” Physical Review D: covering particles, fields, gravitation, and cosmology, vol. 99, Article ID 034505, 2018. View at Publisher · View at Google Scholar
  59. I. Scimemi and A. Vladimirov, “Power corrections and renormalons in transverse momentum distributions,” Journal of High Energy Physics, vol. 2017, no. 2, 2017. View at Publisher · View at Google Scholar
  60. T. Luebbert, J. Oredsson, and M. Stahlhofen, “Rapidity renormalized TMD soft and beam functions at two loops,” High Energy Physics - Phenomenology, vol. 2016, no. 3, p. 168, 2016. View at Google Scholar
  61. J.-y. Chiu, A. Jain, D. Neill, and I. Z. Rothstein, “The rapidity renormalization group,” Physical Review Letters, vol. 108, Article ID 151601, 2012. View at Publisher · View at Google Scholar
  62. T. Gehrmann, T. Luebbert, and L. L. Yang, “Transverse parton distribution functions at next-to-next-to-leading order: the quark-to-quark case,” Physical Review Letters, vol. 109, Article ID 242003, 2012. View at Publisher · View at Google Scholar
  63. M. G. Echevarria, A. Idilbi, and I. Scimemi, “Unified treatment of the QCD evolution of all (un-)polarized transverse momentum dependent functions: Collins function as a study case,” Physical Review D: Particles, Fields, Gravitation and Cosmology, vol. 90, no. 1, Article ID 014003, 2014. View at Publisher · View at Google Scholar
  64. J. C. Collins and A. Metz, “Universality of soft and collinear factors in hard-scattering factorization,” Physical Review Letters, vol. 93, no. 25, Article ID 252001, 2004. View at Publisher · View at Google Scholar
  65. J. R. Gaunt, “Glauber gluons and multiple parton interactions,” Journal of High Energy Physics, vol. 2014, p. 110, 2014. View at Google Scholar
  66. M. Diehl, J. R. Gaunt, D. Ostermeier, P. Pll, and A. Schfer, “Cancellation of Glauber gluon exchange in the double Drell-Yan process,” Journal of High Energy Physics, vol. 2016, p. 076, 2016. View at Publisher · View at Google Scholar
  67. D. Boer, T. van Daal, J. R. Gaunt, T. Kasemets, and P. J. Mulders, “Colour unwound - disentangling colours for azimuthal asymmetries in Drell-Yan scattering,” High Energy Physics - Phenomenology, vol. 3, p. 040, 2017. View at Publisher · View at Google Scholar
  68. S. D. Drell and T.-M. Yan, “Massive lepton-pair production in hadron-hadron collisions at high energies,” Physical Review Letters, vol. 25, p. 316, 1970. View at Google Scholar
  69. G. Altarelli, R. Ellis, and G. Martinelli, “Leptoproduction and drell-yan processes beyond the leading approximation in chromodynamics,” Nuclear Physics B, vol. 143, no. 3, pp. 521–545, 1978. View at Publisher · View at Google Scholar
  70. R. D. Tangerman and P. J. Mulders, “Intrinsic transverse momentum and the polarized Drell-Yan process,” Physical Review D: Particles, Fields, Gravitation and Cosmology, vol. 51, no. 7, pp. 3357–3372, 1995. View at Publisher · View at Google Scholar
  71. G. Kramer and B. Lampe, “Two Jet Cross-Section in e+ e- Annihilation,” Zeitschrift für Physik C Particles and Fields, vol. 34, no. 4, pp. 497–522, 1989. View at Publisher · View at Google Scholar
  72. T. Matsuura, S. C. van der Marck, and W. L. van Neerven, “The calculation of the second order soft and virtual contributions to the Drell-Yan cross section,” Nuclear Physics B, vol. 319, no. 3, pp. 570–622, 1989. View at Google Scholar
  73. A. Idilbi, X.-d. Ji, and F. Yuan, “Resummation of threshold logarithms in effective field theory for DIS, Drell–Yan and Higgs production,” Nuclear Physics B, vol. 753, no. 1-2, Article ID 0605068, pp. 42–68, 2006. View at Publisher · View at Google Scholar
  74. I. Balitsky and A. Tarasov, “Power corrections to TMD factorization for Z-boson production,” High Energy Physics - Phenomenology, vol. 2018, p. 150, 2018. View at Google Scholar
  75. M. A. Ebert, I. Moult, I. W. Stewart, F. J. Tackmann, G. Vita, and H. X. Zhu, “Power corrections for N-jettiness subtractions at ,” Journal of High Energy Physics, vol. 84, 2018. View at Google Scholar
  76. J. Collins, L. Gamberg, A. Prokudin, T. Rogers, N. Sato, and B. Wang, “Relating transverse-momentum-dependent and collinear factorization theorems in a generalized formalism,” Physical Review D: Particles, Fields, Gravitation and Cosmology, vol. 94, no. 3, Article ID 034014, 2016. View at Publisher · View at Google Scholar
  77. D. Pitonyak, L. Gamberg, A. Metz, and A. Prokudin, “Connections between collinear and transverse-momentum-dependent polarized observables within the Collins-Soper-Sterman formalism,” Physics Letters B, vol. 781, pp. 443–454, 2018. View at Publisher · View at Google Scholar
  78. C. T. H. Davies, B. R. Webber, and W. J. Stirling, “Nucl. Phys,” B256, 413 (1985).
  79. R. K. Ellis and S. Veseli, “W and Z transverse momentum distributions: resummation in qT-space,” Nuclear Physics B, vol. 511, p. 649, 1998. View at Publisher · View at Google Scholar
  80. G. P. Korchemsky and G. F. Sterman, “Nonperturbative corrections in resummed cross sections,” Nuclear Physics B, vol. 437, no. 2, pp. 415–432, 1995. View at Publisher · View at Google Scholar
  81. J. C. Collins and D. E. Soper, “Back-to-back jets in QCD,” Nuclear Physics B, vol. 193, no. 2, pp. 381–443, 1981. View at Google Scholar
  82. Y. Li, D. Neill, and H. X. Zhu, “An exponential regulator for rapidity divergences,” Physical Review D: Particles, Fields, Gravitation and Cosmology, 2016. View at Google Scholar
  83. A. Bacchetta, F. Delcarro, C. Pisano, M. Radici, and A. Signori, “Extraction of partonic transverse momentum distributions from semi-inclusive deep-inelastic scattering, Drell-Yan and Z-boson production,” Journal of High Energy Physics, vol. 2017, p. 81, 2017. View at Google Scholar
  84. T. Becher and G. Bell, “Enhanced Nonperturbative Effects through the Collinear Anomaly,” Physical Review Letters, vol. 112, no. 18, 2014. View at Publisher · View at Google Scholar
  85. V. Bertone, I. Scimemi, and A. Vladimirov, “Extraction of unpolarized quark transverse momentum dependent parton distributions from Drell-Yan/Z-boson production,” High Energy Physics - Phenomenology, 2019. View at Google Scholar
  86. P. M. Nadolsky, D. R. Stump, and C. P. Yuan, “Phenomenology of multiple parton radiation in semi-inclusive deep-inelastic scattering,” Physical Review D: Covering Particles, Fields, Gravitation, and Cosmology, vol. 64, Article ID 114011, 2001. View at Google Scholar
  87. G. Bozzi, S. Catani, G. Ferrera, D. de Florian, and M. Grazzini, “Production of Drell–Yan lepton pairs in hadron collisions: Transverse-momentum resummation at next-to-next-to-leading logarithmic accuracy,” Nuclear Physics B, vol. 696, no. 207, pp. 207–2013, 2011. View at Google Scholar
  88. Z.-B. Kang, J.-W. Qiu, X.-N. Wang, and H. Xing, “Next-to-leading order transverse momentum broadening for Drell-Yan production in P + A collisions,” Physical Review D: Covering Particles, Fields, Gravitation, and Cosmology, vol. 94, Article ID 074038, 2016. View at Google Scholar
  89. Z.-B. Kang, X. Liu, F. Ringer, H. Xing, and JHEP., “The transverse momentum distribution of hadrons within jets,” Journal of High Energy Physics, vol. 2017, p. 68, 2017. View at Google Scholar
  90. Z.-B. Kang, A. Prokudin, F. Ringer, and F. Yuan, “Collins azimuthal asymmetries of hadron production inside jets,” Physics Letters B, vol. 2017, pp. 635–642, 2017. View at Publisher · View at Google Scholar
  91. D. Gutierrez-Reyes, I. Scimemi, W. J. Waalewijn, and L. Zoppi, “Transverse-momentum-dependent distributions with jets,” Physical Review Letters, vol. 121, no. 16, 2018. View at Publisher · View at Google Scholar
  92. X. Liu, F. Ringer, W. Vogelsang, and F. Yuan, “Lepton-jet correlations in deep inelastic scattering at the electron-ion collider,” High Energy Physics - Phenomenology, 2018. View at Google Scholar
  93. M. G. A. Bu_ng, Z.-B. Kang, K. Lee, and X. Liu, “A transverse momentum dependent framework for back-to-back photon+jet production,” High Energy Physics - Phenomenology, 2018. View at Google Scholar