Momentum imbalance observables as a probe of gluon TMDs

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Cristian Pisano∗

Department of Physics, University of Antwerp, Groenenborgerlaan 171, 2020 Antwerp, Belgium E-mail:cristian.pisano@uantwerp.be

The unpolarized and linearly polarized gluon TMDs can be directly probed in heavy quark and jet pair production in unpolarized electron-proton collisions by looking at observables, like trans-verse momentum distributions and azimuthal asymmetries, depending on the momentum imbal-ance of the pair. Analytical expressions are presented for these observables and for analogous ones in Higgs plus jet and quarkonium plus photon production in unpolarized proton-proton scat-tering experiments. It is shown how the proposed measurements, to be performed at a future EIC and at the LHC, could provide important information on the size and shape of gluon TMDs, as well as on other fundamental properties such as their process and energy scale dependences.

QCD Evolution 2015 -QCDEV2015-26-30 May 2015

Jefferson Lab (JLAB), Newport News Virginia, USA

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1. Introduction

Transverse momentum dependent parton distribution functions (TMDs) give valuable insight into the intrinsic motion of partons and the correlation between their spins and momenta, providing a full three-dimensional picture of hadrons in momentum space. Their knowledge is necessary to understand the observed polarization effects in high-energy scattering processes, like transverse

single spin asymmetries [1, 2, 3], which have gathered considerable attention from both

experi-mental and theoretical communities in recent years. From the theory point of view, proofs of TMD factorization exist for a few processes in which two energy scales are present: a hard one, Q, and a transverse momentum, qT, such that qT Q. Typical examples are semi-inclusive deeply inelastic scattering (SIDIS), Drell-Yan (DY) and e+e−annihilation processes [4,5,6,7]. On the experimen-tal side, much effort has been made to extract the quark TMDs from low energy data, mainly from HERMES, COMPASS and JLab measurements [1,2]. On the contrary, almost nothing is known experimentally about gluon TMDs [8,9].

If the proton is unpolarized, in addition to the unpolarized gluon TMD denoted by f₁g, a dis-tribution of linearly polarized gluons, h⊥ g₁ , corresponding to an interference between +1 and −1 gluon helicity states, can in principle be nonzero [10]. In Ref. [11] it was found that h⊥ g₁ can contribute to the dijet imbalance in hadronic collisions, commonly used to determine the average transverse momentum squared of partons inside protons. Depending on its size and on whether its contribution can be calculated and taken into account, h⊥ g₁ can complicate or even hamper this ex-traction. Hence it is important to determine its size separately, using other observables. Although in Ref. [11] a method has been discussed to isolate h⊥ g₁ by means of an azimuthal angular dependent weighting of the cross section, such a process is expected to suffer from contributions that break factorization, through initial (ISI) and final state interactions (FSI) [12].

In the present contribution, after providing a formal definition of gluon TMDs in QCD, heavy quark pair and dijet production in electron-proton collisions is discussed. Such processes, that could be studied at a future Electron Ion Collider (EIC), offer a theoretically clean and safe way to probe both f₁gand h⊥ g₁ because TMD factorization is expected to hold [13,14] . Furthermore, it is shown how complementary information could come, in proton-proton collisions, from the analysis of Higgs boson production in association with a jet [15] and from quarkonium production in association with a photon [16]. For all these process, estimates are provided for observables sensitive to gluon TMDs, which depend on the transverse momentum imbalance of two particles or jets in the final state.

2. Operator definition of gluon TMDs

The information on the gluon TMDs of an unpolarized proton carrying four-momentum P is encoded in the transverse momentum dependent correlator, defined as follows [10]

Γ[U ] µνg (x, pppT) = n_ρn_σ (p·n)2 Z _{d(ξ ·P) d}2 ξT (2π)3 e ip·ξ_{hP| Tr} c Fµ ρ(0)U[0,ξ ]Fν σ(ξ )U[ξ ,0]0 |Pi LF = − 1 2x gµ νT f g[U ] 1 (x, ppp2T) − pµ TpνT M2 p + gTµ ν p p p2_T 2M2 p h⊥ g [U]₁ (x, ppp2_T) , (2.1)

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where n is a lightlike vector conjugate to P, p is the gluon momentum decomposed as p = xP +

pT+ p−n(with p2T = −ppp

2

T) and F

µ ν _{denotes the gluon field strength. Moreover, a trace over color}

is taken in the integrand, where non-locality is limited to the lightfront (LF), ξ · n = 0. The process

dependent gauge links U_{[0,ξ ]} and U_{[ξ ,0]}0 are path ordered exponentials in the triplet representation

and are needed to ensure the gauge invariance of the correlator.

Since the above correlator cannot be calculated from first principles, its expansion in terms of TMDs is commonly used for phenomenological applications. This expansion is given at the leading

twist level, with the naming convention of Ref. [17], in the second line of Eq. (2.1), where we have

introduced the transverse projector gµ νT = gµ ν− Pµnν/P · n − nµPν/P · n. The function f g[U ] 1 is the distribution of unpolarized gluons inside an unpolarized proton, while h⊥ g [U]₁ represents the T -even distribution of linearly polarized gluons inside an unpolarized proton. The function h⊥ g [U]₁ is also even in pT because it enters in the correlator as a rank 2 tensor, describing a ∆L = 2 helicity-flip distribution. Furthermore, it satisfies the following, model-independent, positivity bound,

|ppp2_T| 2M2 p

|h⊥ g [U]₁_T (x, ppp2_T)| ≤ f₁g[U ](x, ppp2_T) . (2.2)

Like all other TMDs, f₁g[U ]and h⊥ g [U]₁ receive contributions from ISI or FSI, summed up into the gauge links, that lead to their process dependence or even to violations of QCD factorization, as already mentioned above. It is therefore important to consider extractions from several different processes, like the ones proposed in the following. In this way, the nonuniversality and factorization breaking issues can be properly studied and quantified.

3. Heavy quark pair and dijet production in DIS

Assuming TMD factorization, the cross section for the process e(`) + p(P) → e(`0) + Q(K1) + ¯

Q(K2) + X , with the four-momenta of the particles given inside brackets, can be written as

dσ = 1 2s d3`0 (2π)3_2E0 e d3KQ (2π)3_2E Q d3KQ¯ (2π)3_2E ¯ Q Z dx d2ppp_T(2π)4δ4(q+p−KQ−KQ¯) ×

_∑

a,b,c 1 Q4 Tr L(`, q) ⊗ Γa(x,pppT) ⊗ |Hγ∗a→b c(q, p, KQ, KQ¯)|2 , (3.1)

where s = (` + P)2, q = ` − `0 is the four-momentum of the exchanged virtual photon with q2= −Q2_{, and the symbol ⊗ denotes appropriate traces over the Lorentz and Dirac indices. The sum} runs over all the partons in the initial and final states, while the leptonic tensor L(`, q) is given by

Lµ ν_{(`, q) = −g}µ ν_Q2_{+ 2 (`}µ_`0ν_{+ `}ν_`0µ_{) ,} _(3.2) and Hγ∗a→bc is the amplitude for the partonic hard subprocess γ

∗_a_{→ bc. At leading order (LO) in} perturbative QCD, only the channel γ∗g→ Q ¯Qcontributes. By substituting the parameterization of the gluon correlator, Eq. (2.1), into Eq. (3.1), one obtains, in the γ∗-P center-of-mass frame and in the specific kinematic configuration in which the quark-antiquark pair is almost back-to-back in

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0 0.05 0.1 0.15 0.2 0.25 0.3 5 10 15 20 25 |K ⊥| R Q2 Q2 Q2 MQ = Mc (GeV) = 100 GeV2 0 0.05 0.1 0.15 0.2 0.25 0.3 5 10 15 20 25 |K ⊥| R Q2 Q2 Q2 MQ = Mc (GeV) = 10 GeV2 0 0.05 0.1 0.15 0.2 0.25 0.3 5 10 15 20 25 |K ⊥| R Q2 Q2 Q2 MQ = Mc (GeV) = 1 GeV2 0 0.05 0.1 0.15 0.2 0.25 0.3 5 10 15 20 25 |K ⊥| R Q2 Q2 Q2 MQ = Mb (GeV) = 100 GeV2 0 0.05 0.1 0.15 0.2 0.25 0.3 5 10 15 20 25 |K ⊥| R Q2 Q2 Q2 MQ = Mb (GeV) = 10 GeV2 0 0.05 0.1 0.15 0.2 0.25 0.3 5 10 15 20 25 |K ⊥| R Q2 Q2 Q2 MQ = Mb (GeV) = 1 GeV2

Figure 1: Upper bounds R on |hcos 2(φ⊥− φT)i| for the process ep → e0Q ¯QX, as a function of |KKK⊥| at

z1= z2= 1/2, y = 0.01 and different values of Q2, for charm (left panel) and bottom (right panel) production.

the plane orthogonal to the direction of P and γ∗,

dσ dyQdyQ¯dy dxxBd2qqqTd2KKK⊥ = α 2 αs π sM_⊥2 1 xBy2 A0+ A1cos φ⊥+ A2cos 2φ⊥ + qqq2_TB0cos 2(φ⊥− φT) + B1cos(φ⊥− 2φT) + B 0 1cos(3φ⊥− 2φT) + B2cos 2φT+ B 0 2cos 2(2φ⊥− φT) δ (1 − z1− z2) . (3.3)

In Eq. (3.3), y_{Q/ ¯}_Q are the rapidities of the final heavy quark/antiquark with mass MQ, z1/2 =

P· K_{Q/ ¯}_Q/P · q, y = P · q/P · ` and xBis the Bjorken variable. Moreover, we have introduced the sum and difference of the transverse heavy quark momenta, K⊥= (KQ⊥− KQ⊥¯ )/2 and qT= KQ⊥+ KQ⊥¯ with |qT| |K⊥|. Therefore, we use the approximate transverse momenta KQ⊥≈ K⊥, KQ⊥¯ ≈ −K⊥, and M_Q⊥2 ≈ M2

¯

Q⊥≈ M⊥2 = MQ2+ KKK 2

⊥. The azimuthal angles of qqqT and KKK⊥, denoted as φT and φ⊥ respectively, are measured w.r.t. the leptonic plane (φ`= φ`0).

The terms Ai in Eq. (3.3) depend only on one TMD, f g 1, while B (0) i depend only on h ⊥ g 1 , and their explicit expressions can be found in Ref. [14]. The different azimuthal modulations in Eq. (3.3) can be singled out by defining the following weighted cross sections

hW (φ⊥, φT)i ≡ R dφ⊥dφTW(φ⊥, φT) dσ R dφ⊥dφTdσ , (3.4)

with W (φ⊥, φT) being one of the circular functions of φ⊥ and φT in Eq. (3.3).

By means of the positivity bound in Eq. (2.2), one can determine the maximum values of the asymmetries originating from the linear polarization of gluons. For example, the upper bound R on |hcos 2(φ⊥− φT)i| is defined as follows:

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0 0.1 0.2 0.3 0.4 0.5 5 10 15 20 25 |K_⊥| R’ Q2 Q2 Q2 MQ = Mc (GeV) = 100 GeV2 0 0.1 0.2 0.3 0.4 0.5 5 10 15 20 25 |K_⊥| R’ Q2 Q2 Q2 MQ = Mc (GeV) = 10 GeV2 0 0.1 0.2 0.3 0.4 0.5 5 10 15 20 25 |K_⊥| R’ Q2 Q2 Q2 MQ = Mc (GeV) = 1 GeV2 0 0.1 0.2 0.3 0.4 0.5 5 10 15 20 25 |K_⊥| R’ Q2 Q2 Q2 MQ = Mb (GeV) = 100 GeV2 0 0.1 0.2 0.3 0.4 0.5 5 10 15 20 25 |K_⊥| R’ Q2 Q2 Q2 MQ = Mb (GeV) = 10 GeV2 0 0.1 0.2 0.3 0.4 0.5 5 10 15 20 25 |K_⊥| R’ Q2 Q2 Q2 MQ = Mb (GeV) = 1 GeV2

Figure 2: Same as in Fig. 1, but for the upper bound R0on cos 2φT

M_c2= 2 GeV2_{, M}2

b= 25 GeV2. Similarly, the upper bound R

0_{on |hcos 2φ} Ti| is given by |hcos 2φTi| = R dφ⊥dφT cos 2φTdσ R dφ⊥dφTdσ =qqq 2 T|B2| 2 A0 = qqq 2 T 2M2 |h⊥ g₁ (x, ppp2_T)| f₁g(x, ppp2 T) R0 ≤ R0, (3.6)

and shown in Fig. 2 in the same kinematic region as Fig.1. Note that R0 becomes larger than

Ronly at small values of |KKK⊥|, but falls off more rapidly than R as |KKK⊥| increases. The size of these asymmetries, together with the relative simplicity of the proposed measurement that does not require polarized beams, will likely allow an extraction of h⊥ g₁ at a future EIC. On the other hand, the unpolarized gluon TMD could be easily determined by integrating Eq. (3.3) over φ⊥ and φT.

The cross section for dijet production has the same structure as the one in Eq. (3.3), although in this case the Aiterms will depend on the partonic process γ∗q→ qg as well. Hence, the upper bounds on the azimuthal asymmetries can be obtained from those for heavy quark pair production in the limit MQ→ 0.

4. Higgs plus jet production at the LHC

In addition to the EIC, gluon TMDs can be studied at the LHC as well. For example, inclusive Higgs production has been analyzed within the TMD factorization approach in Refs. [18,19] and, including the effects of QCD evolution, in Refs. [20,21,22]. It turns out that the impact of gluon polarization on the Higgs transverse momentum distribution is of the order of a few percent, at the energy scale of the Higgs mass MH. Furthermore, such effects are largest when the transverse momentum of the Higgs boson is small, namely a few GeV, i.e. in a region where the cross section is difficult to measure. In this section the associated production of a Higgs boson and a jet [15] is discussed. This process offers some additional features compared to inclusive Higgs production, when it comes to probe gluon TMDs. First of all, one can study their scale evolution by tuning the hard scale, identified for example with the invariant mass of the Higgs-jet pair. This is not possible in Higgs production, for which the hard scale is fixed to be MH. Moreover, one can define angular modulations that allow to select specific convolutions of TMDs. Finally, h⊥ g₁ affects the spectrum of the Higgs-jet pair mostly at values of the transverse momentum of the pair as small as a few GeV, but the single transverse momenta of the Higgs boson and the jet can be much larger.

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To LO in perturbative QCD, the reaction p(PA) + p(PB) → H(KH) + jet(Kj) + X , with the

Higgs boson and the jet almost back to back in the plane orthogonal to the direction of the

ini-tial protons, proceeds via the following partonic subprocesses: gg → Hg, gq → Hq and q ¯q→ Hg.

Consider only the channel gg → Hg, which is the dominant one at the LHC energies, and assuming

TMD factorization, the cross section is given by, , similarly to Eq. (3.1),

dσ = 1 2s d3KKKH (2π)3_2E H d3KKKj (2π)3_2E j Z dxadxbd2pppaTd 2_p_p_p bT(2π) 4 δ4(pa+pb−KH− Kj) ×TrnΓg(xa,pppaT) ⊗ Γg(xb,pppbT) ⊗ Hgg→Hg(pa, pb, KH, Kj) 2o , (4.1)

where s = (PA+ PB)2. By neglecting terms suppressed by powers of |qqqT|/M⊥, with qqqT= KKKH⊥+ KKKj⊥

and M⊥=

q

M_H2+ KKK2_H⊥, the final result in the laboratory frame reads

dσ σ = 1 2π σ0(qqq 2 T)1 + R0(qqq 2 T) + R2(qqq 2 T) cos 2φ + R4(qqq 2 T) cos 4φ , (4.2)

where the normalized cross section is given by dσ σ ≡ dσ Rq2_Tmax 0 dqqq2T R2π 0 dφ dσ , with dσ ≡ dσ dyHdyjd2KKK⊥d2qqqT . (4.3)

In the definition above, yHand yj are, respectively, the rapidities of the produced Higgs boson and

jet along the direction of the incoming protons, KKK⊥= (KKKH− KKKj)/2 ≈ KKKH⊥≈ −KKKj⊥, and φ is the

azimuthal angle between KKK⊥and qqqT. Moreover, we have introduced

σ0(qqq2T) = C [ fg 1 f g 1] Rq2_Tmax 0 dqqq2TC [ f g 1 f g 1] , (4.4)

and the convolution of TMDs

C [w f g] ≡Z d2ppp_a_T Z d2ppp_b_Tδ2(pppaT+ pppbT− qqqT) w(pppaT, pppbT) f (xa, ppp 2 aT) g(xb, ppp 2 bT) , (4.5)

with xa/b = M⊥e±yH + |KKKj⊥| e±yj /

√

s, up to corrections of orderO(1/s). The terms R0, R2, R4

in Eq. (4.2) are functions of qqq2_Tas well as of the Mandelstam variables for the subprocess gg → Hg:

R₀(qqq2_T) = M 4 Hsˆ2 M_H8+ ˆs4_{+ ˆt}4_{+ ˆ}_u4 C [whh 0 h ⊥ g 1 h ⊥ g 1 ] C [ fg 1 f g 1] , R2(qqq2T) = ˆt2_{(ˆt+ ˆ}_u)2_{− 2M}2 Huˆ2(ˆt+ û) + MH4(ˆt2+ û2) M_H8+ ˆs4_{+ ˆt}4_{+ ˆ}_u4 C [wf h 2 f g 1h ⊥ g 1 ] C [ fg 1 f g 1] + (xa↔ xb, ˆt ↔ û) , R₄(qqq2_T) = ˆt 2_u_ˆ2 M_H8+ ˆs4_{+ ˆt}4_{+ ˆ}_u4 C [whh 4 h ⊥ g 1 h ⊥ g 1 ] C [ fg 1 f g 1] , (4.6)

where the transverse weights are explicitly given by

whh₀ = 1 M4 p (ppp_a_T· ppp_b_T)2−1 2ppp 2 aTppp 2 bT , w₂f h = 1 M2 p 2(qqqT· pppbT) 2 q q q2 T − ppp2_b_T , wh f₂ = 1 M2 p 2(qqqT· pppaT) 2 q q q2 T − ppp2_a_T , whh₄ = 1 2M4 p ( 2 2(qqqT· pppaT)(qqqT· pppbT) q q q2 T − ppp_a_T· ppp_b_T 2 − ppp2_a_Tppp2_b_T ) . (4.7)

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0 0.02 0.04 0.06 0.08 0.1 0 0.5 1 1.5 2 q_T σ-1dσ/d

Gaussian + tail, max pol.

q_T2 [GeV] [GeV-2] 0 0.02 0.04 0.06 0.08 0.1 0 0.5 1 1.5 2 q_T σ-1dσ/d

q_T2 [GeV] [GeV-2] K⊥ = 10 GeV 0 0.02 0.04 0.06 0.08 0.1 0 0.5 1 1.5 2 q_T σ-1dσ/d

q_T2 [GeV] [GeV-2] K⊥ = 100 GeV 0 0.02 0.04 0.06 0.08 0.1 0 0.5 1 1.5 2 q_T σ-1dσ/d

q_T2 [GeV] [GeV-2] lin. pol. = 0 0 0.0005 0.001 0.0015 0.002 0.0025 0.003 0 2 4 6 8 10 12 q_T 〈 cos2φ〉q_T [GeV -2 ]

[GeV] 0 0.0005 0.001 0.0015 0.002 0.0025 0.003 0 2 4 6 8 10 12 q_T 〈 cos2φ〉q_T [GeV -2 ]

[GeV] K⊥ = 10 GeV 0 0.0005 0.001 0.0015 0.002 0.0025 0.003 0 2 4 6 8 10 12 q_T 〈 cos2φ〉q_T [GeV -2 ]

[GeV] K⊥ = 100 GeV 5⋅10-5 0 0.0001 0.00015 0.0002 0.00025 0.0003 0 2 4 6 8 10 12 q_T 〈 cos4φ〉q_T [GeV -2 ]

[GeV] 5⋅10-5 0 0.0001 0.00015 0.0002 0.00025 0.0003 0 2 4 6 8 10 12 q_T 〈 cos4φ〉q_T [GeV -2 ]

[GeV] K⊥ = 10 GeV 5⋅10-5 0 0.0001 0.00015 0.0002 0.00025 0.0003 0 2 4 6 8 10 12 q_T 〈 cos4φ〉q_T [GeV -2 ]

[GeV]

K⊥ = 100 GeV

Figure 3: Transverse momentum distribution (left panel), cos 2φ and cos 4φ azimuthal asymmetries (central and right panels) of the Higgs boson plus jet pair in the process p p → H jet X for two different choices of K⊥, with qqq2Tmax= MH2/4 and yH= yj. The solid line indicates the H+jet spectrum in absence of linear

polarization. The shaded blue regions represent the range of the three observables as K⊥varies from 0 to ∞.

By defining the observables [16]

hcos nφ iqT ≡ R_2π

0 dφ cos nφ dσ

σ , n= 0, 2, 4 , (4.8)

such that their integrals over qqq2_T give the average values of cos nφ , with n = 0, 2, 4,

hcos nφ i ≡ Rq2_T_max 0 dqqq2T R2π 0 dφ cos nφ dσ σ = Z q2 Tmax 0 dqqq2_Thcos nφ iqT, (4.9) it is possible to isolate the amplitudes 1 + R0, R2, R4 in Eq. (4.2) and access the three different convolutions of gluon TMDs. One obtains

1 σ dσ dqqq2 T ≡ h1iqT = σ0(qqq 2 T) [1 + R0(qqq 2 T)] , (4.10) hcos 2φ iqT = 1 2σ0(qqq 2 T) R2(qqq 2 T) , (4.11) hcos 4φ iqT = 1 2σ0(qqq 2 T) R4(qqq 2 T). (4.12)

Predictions for these transverse momentum dependent quantities are presented in Fig. 3 in the

specific configuration in which the Higgs boson and the jet have the same rapidities, yH = yj, and for two different values of K⊥≡ |KKK⊥|, namely 10 and 100 GeV. The so far unknown unpolarized gluon TMD has been assumed to have the following form [21],

f₁g(x, ppp2_T) = f₁g(x)R 2 2 π 1 1 + ppp2 TR2 , (4.13)

where R = 2 GeV−1 and f₁g(x) is the unpolarized gluon distribution, integrated over its transverse momentum. In order to show the maximal effects of gluon polarization, h⊥ g₁ is taken to be positive and saturating the bound in Eq. (2.2). Moreover, qqq2_T_max= M_H2/4 [15]. Note that hcos 2φ iqT is the

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only observable sensitive to the sign of h⊥ g₁ , and it is expected to be negative if h⊥ g₁ > 0. It is found

that |hcos 2φ i| ≈ 12% at K⊥ = 100 GeV, while it is about 0.5% at K⊥ = 10 GeV. Furthermore,

hcos 4φ i ≈ 0.2% if K⊥= 100 GeV and completely negligible at K⊥= 10 GeV. Analogous results

are obtained if one considers a Gaussian model for f₁gand maximal gluon polarization [15].

To conclude this section, it is important to mention that the proposed measurements require high resolution of the transverse momentum of both the Higgs boson and the jet, because one would need several bins in qT in the kinematic region qT ≤ 10 GeV. In addition, the knowledge of how well the jet axis coincides with the direction of the fragmenting parton is needed.

5. Quarkonium production in association with a photon in pp collisions

The calculation of the cross section for the process p(PA) + p(PB) →Q(KQ) + γ(Kγ) + X , whereQ is a C = − quarkonium (J/ψ or ϒ) with mass M_Q, proceeds along the same lines of Higgs plus jet production presented in the previous section. As before, the imbalance of the quarkonium-photon pair qqq_T = KKK_Q⊥+ KKKγ ⊥is small, but not the individual transverse momenta KKKQ⊥ and KKKγ ⊥. Therefore no forward detector is needed for the proposed measurements, in contrast to the the inclusive production of C = + quarkonia that can also be used to probe gluon TMDs [23,24,25].

The resulting cross section dσ ≡ dσ /dQ dY d2qqq_TdΩ has the following structure [16], dσ =N nF1C h f₁gf₁g i + F2C h w₂f hf₁gh⊥g₁ + xa↔ xb i cos 2φ + F4C h whh₄ h⊥g₁ h⊥g₁ i cos 4φo, (5.1) where Q and Y are the invariant mass and the rapidity of the pair, to be measured, like qqq_T, in the hadronic center-of-mass frame. The solid angle Ω = (θ , φ ) is measured in the so-called Collins-Soper frame [26], defined as the frame in which the final pair is at rest and the ˆxˆz-plane is spanned by PPPAand PPPB, with the ˆx-axis set by their bisector. The normalization factorN is given by

N = 4α2 sα e2Q|R0(0)|2 Q2− M2 Q 3 s Q3_M3 QD , (5.2)

where R0(0) is the radial wave function of the quarkonium evaluated at the origin, eQis the heavy quark charge and

D=(α2_{+ 1)}2_{− (α}2_{− 1)}2_cos2

θ2, (5.3)

with α ≡ Q/M_Q. The convolutions of TMDs are defined in Eq. (4.5), with the light-cone momen-tum fractions xa/b= exp[±Y ] Q/

√

s, the transverse weights are given in Eqs. (4.7) and the F1,2,4 terms read F₁= 1 + 2α2+ 9α4+ (6α4− 2) cos2θ + (α2− 1)2cos4θ , F2= −8 α2sin2θ , F4= 1 2(α 2_{− 1)}2_sin4 θ . (5.4)

Note that in the derivation of Eq. (5.1), the Color Singlet Model [27] for the quarkonium production mechanism has been adopted. Color octet contributions [28], which could potentially lead to a breakdown of TMD factorization, are expected to be small, especially for ϒ − γ production in the kinematic region described in the caption of Fig.4[16].

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2 4 6 8 10qTHGeVL 0.005 0.010 0.020 0.050 0.100 KMR CGC Gaussian 2 4 6 8 10qTHGeVL -0.0010 -0.0008 -0.0006 -0.0004 -0.0002 Set B + max KMR + max CGC Gaussian + max 2 4 6 8 10qTHGeVL 0.0001 0.0002 0.0003 0.0004 Set B + max KMR + max CGC Gaussian + max

Figure 4: The TMD observablesSq(0)T ,S (2)

qT andS

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qT for the process p(PA) + p(PB) → ϒ(PQ) + γ(Pγ) + X

at√s= 14 TeV in the kinematic region defined by Q = 20 GeV, Y = 0, θ = π/2, and xa= xb' 1.4 × 10−3.

As in the previous section, the three terms in Eq. (5.1) can be singled out by means of the

observablesSq(n)T ≡ hcos nφ iqT, with n = 0, 2, 4, defined in Eq. (4.8), where the cross section in the denominator is integrated over φ and qqq2_T, up to a value qqq2_T_max Q2_{. One finds}

S(0) qT = C [ fg 1f g 1] R dqqq2 TC [ f g 1f g 1] , Sq(2)T = F2C [w₂f hf₁gh⊥g₁ + xa↔ xb] 2F1 R dqqq2 TC [ f g 1f g 1] , Sq(4)T = F4C [whh4 h ⊥g 1 h ⊥g 1 ] 2F1 R dqqq2 TC [ f g 1f g 1] . (5.5) Numerical estimates for these observables are presented in Fig. 4 for ϒ + γ production, where q

qq2_T_max= Q2/4. In contrast to Higgs plus photon production, h⊥ g₁ does not contribute to the spectrum of the quarkonium-photon pair. Therefore, through a measurement ofSq(0)T , which turns out to be sizable, one could determine the shape of f₁gas a function of qT. Furthermore, sinceS

(2)

qT andS

(4) qT are rather small, it would be needed to integrate them over qqq2_T, e.g. up to Q2/4, to get at least an experimental evidence of a nonzero linear polarization of gluons [16].

6. Summary and conclusions

The unpolarized and linearly polarized gluon TMDs of an unpolarized proton could be probed in future EIC experiments by looking at the transverse momentum distributions and azimuthal asymmetries of heavy quark and jet pairs. At the LHC complementary information can be gath-ered by similar analyses for processes like pp → H jet X , pp → J/ψ(ϒ) γ X , presented in detail in this contribution, or like pp → γ γ X , proposed in Ref. [29] specifically for RHIC and for which factorization breaking terms should be absent.

Although TMDs are not universal, their process dependence can be calculated. For instance, despite the different gauge link structures, it has been found that in ep → e0Q ¯Q X and in all the processes with a colorless final state, like pp → γ γ X , pp → H X and pp → ηc,bX, one always probes the same effective distribution of linearly polarized gluons, given by the sum of two of the five universal h⊥ g₁ functions [14,30]. This restricted universality needs to be tested experimentally, using LHC or RHIC data. On the other hand, the gauge link structure for pp → H jet X is much more complicated and still has to be investigated. It is therefore of great interest to compare the extractions of gluon TMDs from different processes, in order to learn about their process and energy scale dependences, as well as the size and importance of possible factorization breaking effects.

Acknowledgments

(10)

Jean-PoS(QCDEV2015)024

Philippe Lansberg, Piet Mulders, and Marc Schlegel for fruitful collaborations on this topic. This work was supported by the Fonds Wetenschappelijk Onderzoek - Vlaanderen (FWO) through a Pegasus Marie Curie Fellowship.

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