Charged Higgs Signals in t t ¯ H Searches
Daniele S. M. Alves, Sonia El Hedri, Anna Maria Taki, Neal Weiner
CCharged Higgs Signals in t t H
Searches
Daniele S. M. Alves,
1, 2, ∗ Sonia El Hedri, † Anna Maria Taki, ‡ and Neal Weiner § Center for Cosmology and Particle Physics,Department of Physics, New York University, New York, NY 10003 Department of Physics, Princeton University, Princeton, NJ 08544 PRISMA Cluster of Excellence & Mainz Institute for Theoretical Physics,Johannes Gutenberg University, 55099 Mainz, Germany (Dated: March 28, 2017)
Abstract
New scalars from an extended Higgs sector could have weak scale masses and still have escapeddetection. In a Type I Two Higgs Doublet Model, for instance, even the charged Higgs can belighter than the top quark. Because electroweak production of these scalars is modest, the greatestopportunity for their detection might come from rare top decays. For mass hierarchies of the type m t > m H + > m A , H , the natural signal can arise from top quark pair production, followed by thedecay chain t → bH + , H + → W +( ∗ ) φ , φ → bb, τ + τ − , where φ = A , H . These final stateslargely overlap with those of the Standard Model ttH SM process, and therefore can potentiallycontaminate ttH SM searches. We demonstrate that existing ttH SM analyses can already probe lightextended Higgs sectors, and we derive new constraints from their results. Furthermore, we notethat existing excesses in ttH SM searches can be naturally explained by the contamination of raretop decays to new light Higgses. We discuss how to distinguish this signal from the Standard Modelprocess. ∗ [email protected] † [email protected] ‡ [email protected] § [email protected] a r X i v : . [ h e p - ph ] M a r ONTENTS
I. Introduction 3II. Model and Signals 5A. Charged Higgs Decays 7III. Constraints on a Light Higgs Sector 8A. Indirect Bounds 9B. Collider bounds 9IV. Signals in t ¯ tH SM Searches 11V. Recast of the ATLAS 13 TeV Search in Multileptons 14VI. Discussion 18Acknowledgments 19A. Recasting the ATLAS search for t ¯ tH SM in multileptons 191. Overview 192. The recasting procedure 203. Signal generation 22B. Signal yields and statistical procedure 231. Signal yields 232. Branching ratios 243. Statistical treatment of fits and exclusions 26References 272 . INTRODUCTION So far, LHC searches have not provided conclusive signs of new particles, nor significantdeviations from Standard Model predictions. Generic limits on new colored particles are par-ticularly severe, with squarks and gluinos from Supersymmetry already tightly constrainedif lighter than 1.3 TeV and 1.9 TeV, respectively [1].Exclusion limits on direct production of new electroweak particles, in contrast, have notbeen as dramatic. Mass limits on charginos produced via electroweak interactions, for in-stance, do not extend beyond 400-500 GeV [2]. The reason for this is straightforward: ata p - p machine such as the LHC, electroweak cross sections are simply much smaller thanstrong cross sections. Because of that, the best means to search for new electroweak states isoften in cascade decays of copiously produced colored particles. This approach however relieson the existence of heavier colored particles within the reach of the LHC. As this possibilitybecomes more and more remote, the only realization of this scenario that can be concretelystudied are rare top quark decays, which unfortunately cannot probe new particles heavierthan about 170 GeV.The range below the top quark mass is nonetheless a well motivated region to search forstates beyond the Standard Model (BSM). One possibility of great interest would be raretop decays to charged Higgs bosons, t → b H + . With the tt cross section at the 13 TeV LHCbeing over 800 pb, even small branching ratios Br( t → b H + ) ∼ O (10 − ) would yield a H ± production rate at the O (pb) level or higher. Since direct electroweak production of thesestates ranges from O (30 − Type II
Two Higgs DoubletModels (2HDM) in particular, flavor changing observables, most notably b → sγ , exclude m H ± ∼ <
580 GeV [3, 4], absent cancellations. Flavor bounds are highly model dependent,however. To contrast, in
Type I b → sγ are mild at best, requiringonly that tan β ∼ > (1 . −
2) for m H + < m t . Thus, at least for Type I 2HDM scenarios, thereis a compelling case to consider direct searches for new Higgs states at colliders, in particular3n rare top decays.A critical point to consider here is that, in a Type I 2HDM, one of the Higgs doubletsis fermiophobic. This can drastically alter the phenomenology of the charged Higgs relativeto a Type II scenario. In particular, if a lighter neutral scalar φ (= A , H ) exists, thecharged Higgs will dominantly decay as H + → φ W +( ∗ ) , and the familiar fermionic decays(e.g., H + → τ + ν, c ¯ s, t ∗ b ) will be suppressed. In this region of parameter space, our processof interest will be: p p → t ¯ t → ( bW + )(¯ b H − ) → ( bW + )(¯ b W − ( ∗ ) φ ) (1.1)with φ dominantly decaying as φ → b ¯ b , τ + τ − , (1.2)if m φ ∼ <
110 GeV. The resulting final state, with rates of O (pb) or higher, can lead toa large signal contamination in searches for the SM Higgs boson produced in associationwith top quark pairs, t ¯ tH SM , whose SM cross section is about 0 . t ¯ tH SM cross section can be used to constrain the Type I 2HDMsignatures in (1.1), (1.2). A more exciting prospect would be to explain recent excesses inexisting t ¯ tH SM searches as due to contamination from rare top decays to charged Higgses.While the significance of current excesses is mild, upcoming results with more data will leadto a clearer picture of the excess pattern, would it persist.The layout of this paper is as follows: in Sec. II we briefly review Type I 2HDMs anddescribe the charged Higgs phenomenology in the light mass region. Direct and indirectbounds on the relevant region of parameter space are reviewed in Sec. III. In Secs. IV, V, wediscuss how t ¯ tH SM searches can be used to constrain charged Higgs production, and describethe degree to which the claimed excess in various channels can be explained by this model.Finally, in Sec.VI we discuss the implications of these results and comment on future searchesthat might help better constrain this light region of Type I 2HDMs.4 I. MODEL AND SIGNALS
In a type I 2HDM, one of the Higgs doublets, H , is fermiophobic, and all fermion massesstem from Yukawa couplings to H : L yukawa = H Q Y u U c + H † Q Y d D c + H † L Y ‘ E c + h. c. (2.1)The scalar potential can be generically parameterized as [5]: V scalar = λ (cid:16) | H | − v (cid:17) + λ (cid:16) | H | − v (cid:17) + λ (cid:16) ( | H | − v ) + ( | H | − v ) (cid:17) (2.2)+ λ (cid:16) | H | | H | − | H † H | (cid:17) + λ (cid:16) Re( H † H ) − v v (cid:17) + λ (cid:16) Im( H † H ) (cid:17) , where both doublets, H and H , have hypercharge Y = 1 /
2, and for simplicity we assumethat CP is conserved and all parameters in (2.2) are real.Conventionally, the mass eigenstates of this theory are parameterized by two angles; α ,the mixing angle between the CP-even neutral states, and β , defined as tan β ≡ v /v : H H = v v + 1 √ R α H H + i √ R β G A , (2.3) H ± H ± = R β G ± H ± , (2.4)where R α = cos α sin α − sin α cos α , R β = − sin β cos β cos β sin β . (2.5)Usually, the “SM-like” Higgs (corresponding to the state discovered at the LHC withmass m h = 125 GeV) is the lighter CP-even neutral scalar, H . That does not need tobe the case though, and in principle the SM-like Higgs could be the heavier CP-even scalar, H . Since we are interested in both regimes, we will adopt the following, more genericparameterization [6] H H = v v + 1 √ R β + δ H SM H + i √ R β G A , (2.6)5here R β − δ = sin( β − δ ) − cos( β − δ )cos( β − δ ) sin( β − δ ) . (2.7)Here, δ is a parameter that describes the deviation from the alignment limit. If the SM-likeHiggs corresponds to the lightest CP-even scalar, δ is defined by δ ≡ β − α − π/
2. Conversely,if the SM-like Higgs corresponds to the heaviest CP-even scalar, then δ ≡ β − α . Theadvantage of this parameterization is that δ quantifies the deviation from a SM-like Higgs,and there is no discontinuity in our description of fields as the mass hierarchy changes. Thatis, H SM is always the SM-like state, and H is always the state with suppressed couplings tofermions.In terms of the mass eigenstates, the Yukawa couplings can be re-written as: L yukawa = X f ξ H SM m f v H SM f f c + ξ H m f v H f f c + i ξ f A m f v A f f c + ξ f A m f v √ U ff H ± f f c + h. c. , (2.8)where v = 246 GeV, U ff is a CKM matrix element if f, f c are quarks, and U ff = 1 if f, f c are leptons. Moreover, ξ H SM = cos δ − sin δ tan β , ξ H = sin δ + cos δ tan β , ξ u A = − ξ d,e A = 1tan β . (2.9)Likewise, the EW symmetry breaking couplings of scalars to vector bosons are given by: L φV V = ζ φ m W v φ W + W − + ζ φ m Z v φ ZZ , (2.10)where φ generically denotes H SM , H , A , and, ζ H SM = cos δ , ζ H = sin δ , ζ A = 0 . (2.11)In Type I 2HDMs, unlike Type II, the couplings of A and H ± to fermions are suppressedby tan β , as shown in (2.8) and (2.9), and so are the H yukawa couplings in the alignmentlimit δ → H ± production from top decays, we show in Fig. 1the branching ratio Br( t → b H + ) as a function of m H ± and tan β . Note that even at largetan β ∼ O (10), branching ratios of O (10 − ) are possible, while at low tan β ∼ <
3, the topbranching ratio to H ± can reach the few percent level, Br( t → b H + ) ∼ O (3%) − roughlythe limit where tension might arise with measurements of the top pair cross section σ t ¯ t [7–15].6
00 110 120 130 140 150 160246810 m H + H GeV L t a n b FIG. 1. Contours of Br( t → b H + ) in a Type I 2HDM, as a function of m H ± and tan β . A. Charged Higgs Decays
The decay patterns of the charged Higgs can vary dramatically across the parameter spaceof 2HDMs, significantly impacting the experimental strategies to search for this state.Due to the large popularity of Type II 2HDMs, the most thoroughly explored H ± decays,both in phenomenological studies as well as in experimental efforts, have been the fermionicmodes. This is due to the tan β enhancement of the H ± couplings to down type quarks andleptons, causing the τ ν , cs and b t ∗ modes to dominate the H ± branching ratios. ATLAS[16, 17] and CMS [18, 19] have extensively searched for signatures of t → b ( H + → τ + ν ), andplaced upper bounds on the overall top branching ratio to this final state at the sub-percentlevel. Analogous searches [20, 21] for the H + → c ¯ s, c ¯ b channels have also set percent-levelbounds on the corresponding branching ratios.The same constraints are applicable in Type I 2HDMs if the only kinematically openchannels for H ± decays are light fermions, which is the case if m H ± ∼ < m H , m A . However,if the spectrum contains a lighter neutral scalar with a large H component, such as A or H , the mode H + → W ± ( ∗ ) A /H would naturally dominate the H ± branching ratio, even7
00 110 120 130 140 150 160 1700.010.020.050.100.200.501.00 m H + H GeV L B r H + ! W + ( ⇤ ) A /H H + ! ¯ b t ⇤ H + ! c ¯ s + c ¯ b H + ! ⌧ + ⌫ FIG. 2. Charged Higgs branching ratios in a Type I 2HDM, assuming m A = m H = 100 GeV,tan β = 3, and δ = 0. with the 3-body phase space suppression of an off-shell W ∗ . This is due to the unsuppressedgauge couplings that mediate this decay mode, and the large suppression of the competingfermionic modes, stemming from the smallness of the Yukawa couplings and from the 1 / tan β dependence of the H ± coupling to fermions. This effect is illustrated in Fig. 2, where wehave set m A = m H = 100 GeV, tan β = 3, and δ = 0. The suppression of fermionicmodes is even more pronounced for larger values of tan β , or lower masses of A or H .This qualitatively different phenomenology of charged Higgs decays has been previouslynoted in phenomenological studies [22–28], but to the best of our knowledge, no dedicatedexperimental analysis has explicitly searched for these signatures. A critical question is then:could such a particle have contaminated studies of other SM or BSM processes? And if so,what constraints could existing searches place on these particular charged Higgs signals? III. CONSTRAINTS ON A LIGHT HIGGS SECTOR
While new scalars with sizeable couplings to SM fermions or gauge bosons are subject toconstraints from LEP, Tevatron, and LHC data, additional Higgs bosons with suppressed8ukawa couplings are more elusive to existing searches. In this section, we summarize theconstraints on the charged and neutral Higgs bosons of Type I 2HDMs, with a focus on thelight mass region.
A. Indirect Bounds
As previously mentioned, indirect bounds on light Type I 2HDMs are mild. Besides b → sγ already discussed in Sec. I, other competing constraints from flavor observables are B s → µ + µ − and ∆ M B s,d , which are only marginally stronger, requiring that tan β ∼ > . − . m H ± < m t [3].Another source of indirect constraints comes from contributions to the electroweak obliqueparameters, particularly T , induced by the mass splittings between H ± , A and H . Forthe light spectra considered here, however, we have checked that ∆ T constraints are easilyevaded. B. Collider bounds
LEP placed a robust lower bound on m H ± ∼ > . H + → c ¯ s, τ + ν , assuming the absence of any non-fermionic decays [29]. The DELPHI andOPAL collaborations also considered the bosonic decay H ± → W ±∗ A [30, 31]. In TypeI 2HDM scenarios, the tan β -independent limits obtained were m H ± ∼ > . m H ± ∼ >
65 GeV (OPAL), provided that m A ∼ >
12 GeV.The OPAL collaboration has searched for the associated production e + e − → A H , with A , H → q ¯ q, gg , and τ + τ − [36, 37]. While the resulting mass limits vary across the param-eter space, they essentially turn off for either A or H heavier than ∼
80 GeV.LEP, Tevatron and LHC searches for the SM Higgs are also potentially sensitive to theneutral scalars, A and H . If A , H are lighter than ∼
110 GeV, however, these statesdecay dominantly to b ¯ b final states (with τ + τ − as the subleading mode), and are challengingto probe at the Tevatron and LHC due to the large backgrounds and suppressed crosssections relative to the SM Higgs. While SM Higgs searches at LEP cannot constrain A due9
20 130 140 150 160406080100120140160 m H + H GeV L m A H G e V L FIG. 3. Reinterpretation of the 8 TeV CMS limits on Br( t → b H + ) × Br( H + → τ + ν ) [18] aslower bounds on tan β in a Type I 2HDM, assuming that m H > m H ± . to the absent A Z Z coupling (see Eqs. (2.10,2.11)), they are sensitive to e + e − → H Z production, and constrain ζ H ∼ < . − . m H ’ (15 − φ (= A or H ) is ligherthan m H SM /
2, the decay channel H SM → φ φ can easily dominate the SM Higgs widthfor generic values of the quartic couplings in (2.2). In order to avoid conflict with observa-tions, in the mass range m φ (cid:46)
62 GeV the tri-scalar coupling λ φφH SM must be suppressed, λ φφH SM (cid:46) (2 −
6) GeV. While this condition is not generically satisfied, it is a parame-ter that can be adjusted independently of the physical masses and mixing angles δ and β .Since the charged Higgs phenomenology we will consider is not directly affected by λ φφH SM ,we will include the region m φ (cid:46)
62 GeV in our study, with the implicit assumption thatΓ( H SM → φ φ ) is not in conflict with observations.Another parameter that is directly constrained by SM Higgs measurements is δ – currentdata pushes the model towards the alignment limit where δ is small and the properties of10 (cid:27)(cid:42)(cid:31)(cid:1)(cid:23)(cid:40)(cid:41)(cid:1)(cid:13)(cid:31)(cid:29)(cid:27)(cid:49)(cid:43)(cid:1)(cid:35)(cid:39)(cid:1)(cid:18)(cid:35)(cid:33)(cid:34)(cid:44)(cid:1)(cid:23)(cid:49)(cid:41)(cid:31)(cid:5)(cid:17)(cid:1)(cid:7)(cid:16)(cid:13)(cid:19) t b W +( ⇤ ) bb, ⌧ ⌧ t b W (cid:22)(cid:27)(cid:38)(cid:31)(cid:1)(cid:32)(cid:35)(cid:39)(cid:27)(cid:37)(cid:1)(cid:43)(cid:44)(cid:27)(cid:44)(cid:31)(cid:1)(cid:27)(cid:43)(cid:1) ttH A , H H + FIG. 4. Signal from rare top decay to b H ± in a light Type I 2HDM, whose final states overlapwith those of SM t ¯ tH SM . the 125 GeV Higgs are “SM-like”. For a more thorough discussion on that, see [6, 33–35].Previously mentioned upper bounds on Br( t → b ( H + → τ + ν )) from ATLAS and CMSare sensitive enough to be relevant even if the decay mode H + → τ + ν is subdominant, asis the case of the Type I 2HDMs we are considering. We recast the 8 TeV CMS constraintson the branching ratio Br( t → b H + ) × Br( H + → τ + ν ) [18] as lower bounds on the valueof tan β , displayed in Fig. 3 as a function of m H ± and m A , assuming for simplicity (andwithout loss of generality) that m H > m H ± .To summarize, the light mass region of Type I 2HDMs is still experimentally viable ina vast swath of parameter space. In what follows, we investigate how this region can beconstrained by existing LHC searches for the Standard Model t ¯ t H SM process. IV. SIGNALS IN t ¯ tH SM SEARCHES
Although the parameter space of Type I 2HDMs with mass hierarchy: M φ = A or H < M H , M H ± < M t (4.1)is still experimentally viable, its phenomenology remains relatively unexplored in comparisonto that of heavier 2HDM spectra. A stricking signal of models with (4.1) are rare top decaysthat can contaminate t ¯ tH SM searches, particularly those targeting leptonic or b ¯ b decays of11he SM Higgs, as illustrated in Fig. 4. Given the very large top pair cross section, andthe fact that Br( t → bH + ) can be as high as a few percent, this contamination can leadto observable excesses in t ¯ tH SM measurements relative to the SM expectation. The excesspattern, however, would appear inconsistent across different channels if interpreted as anenhanced t ¯ tH SM signal strength. Generically, no excess should appear on γγ channels, sincesearches typically require m γγ ≈
125 GeV (within resolution) to specifically target the SMHiggs. On the other hand, one would expect excesses in channels targeting b ¯ b and τ + τ − finalstates, albeit with different strengths. Many of these analyses use Multivariate discriminants(MVAs), such as boosted decisions trees, neural networks, etc., which may be tuned to thespecific final state kinematics of t ¯ tH SM . In those cases, the contamination from rare topdecays may be partially filtered out, depending on how well the MVAs can discriminatebetween the two processes. Normally, details of MVA based studies are not public, and theextraction of limits on contaminant signals is unfeasible.In fact, SM Higgs searches as early as the Tevatron’s could have been contaminatedby rare top decays. The CDF collaboration has searched for t ¯ t ( h → b ¯ b ) over the massrange m h = (100 − O (2 σ ) excess above expectations at m h ∼ (100 − µ t ¯ th = 7 . +4 . − . for m h = 100 GeV, and µ t ¯ th = 8 . +4 . − . for m h = 105 GeV. This would correspond to a rateof roughly (65 ±
40) fb, or, in terms of the inclusive top pair cross section, (0 . − . × σ t ¯ t .This analysis was MVA based, and its mass resolution was limited due to the presence offour b -quarks in the signal final state, leading to a combinatoric ambiguity in identifyingthe b -jets originating from h decays, and therefore to a broadening of the expected m b ¯ b peak. All these factors preclude us from inferring any concrete implications regarding apotential contamination from BSM processes, but the results are nonetheless intriguing,and, if corroborated with further deviations at the LHC, could warrant a re-analysis of theTevatron’s data.At the LHC, existing t ¯ tH SM results from ATLAS and CMS do seem to suggest a pattern ofexcesses inconsistent with the hypothesis of enhanced t ¯ tH SM production, although with smallstatistical significance. At 8 TeV, the uncertainties in t ¯ tH SM measurements are too large to12ffer any indication of an excess or lack thereof ∗ , with a combined best fit of µ ttH = 2 . +0 . − . (ATLAS + CMS, all channels) [41]. Existing 13 TeV data is inconclusive as well, showing noexcess in CMS b ¯ b (MVA) [42] and ATLAS γγ [43], and O (1 σ ) excesses in CMS multileptons(MVA) [44, 45], CMS γγ [46], ATLAS b ¯ b (MVA) [47], and ATLAS multileptons [48]. Of all13 TeV searches to date, only the latter, ATLAS multileptons, employs a traditional cut-and-count procedure † , and therefore is amenable to recasting in terms of our charged Higgssignal. We shall do so in the following section.We end by commenting on our choice of spectrum when reinterpreting the ATLAS multi-lepton results. As previously discussed, if (4.1) is realized, the charged Higgs will dominantlydecay to W ( ∗ ) φ , where φ is the lightest neutral scalar. Since LHC measurements of the125 GeV Higgs push this model into the alignment limit, (sin δ ) ∼ < .
1, the signals in Fig. 4will be essentially independent of whether φ = A or H , sinceBr( H ± → W ± ( ∗ ) H )Br( H ± → W ± ( ∗ ) A ) (cid:12)(cid:12)(cid:12)(cid:12)(cid:12)(cid:12) m A = m H = 1 − (sin δ ) . (4.2)Moreover, if lighter than ∼
110 GeV, A and H will have the same leading branching ratios,namely, Br( φ → bb ) ≈ . φ → τ + τ − ) ≈ .
08. For practical purposes, therefore,we will choose A as the lightest neutral scalar and decouple H (i.e., set m H > m H ± ) forthe remainder of the paper. None of the results that follow change in any significant way if A → H . The only loss of generality that comes with this assumption is the possibility oftwo independent decay modes. However, [48] does not rely on a mass peak, the efficienciesdo not vary dramatically over the range of m A , H considered, and, in the presence of amass hierarchy between A and H , the lighter mode dominates the bosonic decays of H ± .Consequently, even this complication should not impact our results significantly. ∗ One notable exception is the CMS measurement in the same-sign dilepton channel [40], which gives thefollowing best fit for the signal strength: µ ttH = 5 . +2 . − . . † The 4 ‘ category in the CMS multilepton analysis [45] employs a cut-and-count procedure as well, but thefinal measurement has uncertainties substantially larger than the ones in ATLAS, and therefore, is lesssensitive. . RECAST OF THE ATLAS 13 TEV SEARCH IN MULTILEPTONS At the time of writing of this paper, [48] was the most recent ATLAS search for t ¯ tH SM production in the multilepton channel, corresponding to 13 . − of 13 TeV data. It targetedleptonic decays of the SM Higgs in W W ∗ , τ + τ − , and ZZ ∗ , by looking into four exclusivesignal regions, namely, same-sign dileptons and one hadronically decaying τ (2 ‘ τ had ), same-sign dileptons vetoing hadronically decaying τ ’s (2 ‘ τ had ), 3 leptons (3 ‘ ), and 4 leptons (4 ‘ ).Rare top decays in Fig. 4, with A → τ + τ − , can contaminate all of these signal regions. Weperformed a Monte Carlo (MC) study to obtain a quantitative estimate of this contaminationand the sensitivity of [48] to the Type I 2HDM spectra of (4.1). Overall, we were able tovalidate our MC simulation of [48] by reproducing the t ¯ tH SM efficiencies provided by ATLASto within a factor of 2 (12 independent efficiencies in total). This translates into a factor of ∼ ∼
50% uncertainty in our resultsfor tan β (since our signal scales as tan β − ). For a detailed description of our MC study,as well as our statistical treatment of fits and exclusions, we refer the reader to Appxs. A,B. Throughout our study, we include the contribution of the SM t ¯ tH SM process, assuming µ t ¯ tH = 1.We first obtain upper bounds on the contamination from t → bH + , so that the expectednumber of events in the signal regions of [48] are compatible with observations. We interpretthese constraints as a lower bound on tan β in the m H + , m A mass plane, shown in Fig. 5(a).Our signal yield is suppressed in the compressed regions m H ± − m A ∼ <
30 GeV (whereBr( H ± → W ± ( ∗ ) A ) is small), and m t − m H ± ∼ <
20 GeV (where Br( t → b H + ) is small). Inthese regions of suppressed signal yield, the inferred lower bounds on tan β are innocuous.These limits would be stronger but for the presence of excesses in the data. Consequently,it behooves us to see whether we can understand these excesses as arising from rare topdecays. In Fig. 5(b) we show the best fit for tan β at each point in the parameter space of m H + and m A . Notably, the efficiencies for the process t → b H + → bW + ( A → τ + τ − )(before folding in branching ratios) do not vary dramatically across the m H + , m A massplane.In Fig. 6, we show the preferred regions of parameter space, defined from the goodness14 Lower bounds on tan 𝛽 (95% C.L.)
120 130 140 150 160 170406080100120140160 m H + H GeV L m A H G e V L Best fit values for tan 𝛽
120 130 140 150 160 170406080100120140160 m H + H GeV L m A H G e V L FIG. 5. Recast of the ATLAS t ¯ tH SM search in multileptons. (Left) Lower bounds on tan β inferredfrom limits on t → b H + → b W + ∗ ( A → τ + τ − ). (Right) Values of tan β that best fit the data. of fit to the data in [48] . Here, the values of tan β are profiled at each mass point to yieldthe best fit (see Fig. 5(b)). Under this assumption, some regions of parameter space areexcluded by other measuments. In particular, the compressed region m H ± − m A ∼ <
40 GeVis in tension with the CMS bounds on Br( t → b ( H + → τ + ν )). Likewise, the compressedregion m t − m H ± ∼ <
10 GeV is in tension with b flavor observables discussed in Sec. III A.Finally, we select two benchmark points to illustrate the pattern of contamination acrossthe four signal regions of [48]. The first benchmark point, B1, is at low mass, m H ± =130 GeV, m A = 40 GeV. For this benchmark, H ± can decay to an on-shell W ± , yieldinga harder charged lepton. Direct production of H + H − and H ± A is several hundreds offemtobarns at 13 TeV, and in principle dedicated searches in final states with 1 or 2 leptonsplus 3 or more b -jets could be sensitive to this point.The second benchmark, B2, is at high mass, and close to the excluded compressed regions, m H ± = 160 GeV, m A = 95 GeV. Because of the small mass splitting between t and H + , the b -jet from t → b H + is relatively soft and has a lower probability of passing the b -taggingrequirements. Overall, B2 predicts that roughly 72% of its signal in the 2 ‘ τ had region has15
20 130 140 150 160 170406080100120140160 m H + H GeV L m A H G e V L
90% C.L.
Preferred regions by the ATLAS excess
68% C.L. flavor observables Br t ! b ( H + ! ⌧ + ⌫ ) benchmark B1benchmark B2 FIG. 6. Preferred regions from the fit to the data in [48], at 68% C.L. (green region), and 90% C.L.(yellow region). Also shown: regions in tension with indirect flavor observables, and in tension withCMS limits on Br( t → b ( H + → τ + ν )). These tensions only apply if the best tan β fit is assumedfor each mass point. only one b -tagged jet. Indeed, the data in the 2 ‘ τ had channel seem to indicate that theexcess appears exclusively in the 1 b -tag category. Since current statistics is very low, thismight be due to Poisson fluctuations, and only more data will be able to settle this matter.A final comment on B2 regards our choice of m A for this benchmark. This was motivatedby an excess at m H ≈ (90 − H , if this were the lightestneutral scalar with ζ H ∼ O (10 − ) (see Eqs.(2.10),(2.11)).In Fig. 7, we show the predicted signal strength from benchmarks B1 and B2, measuredin units of the SM t ¯ tH SM signal strength, for each of the four signal regions of [48]. Boththe combined best fit, as well as a selected range of tan β are shown. Specifically, this rangeis chosen so it encompasses the 1 σ range favored by data in the 2 ‘ τ had channel. We pointout that for B2, it is not possible to reach the 1 σ upper range of 2 ‘ τ had without violating16 ` ⌧ had ` ⌧ had ` ` µ t ¯ tH (signal strength relative to SM ) t ¯ tH SM Ê ÊÊ » »»»» »» << Ê ÊÊÊ » »»» » »»»
Ê ÊÊ Ê » »» »» »» » - ATLAS 2 𝜎 upper limitB1 combined best fit »» »» ÊÊ B1 varying over range ` ⌧ had B2 combined best fit »» »»
B2 varying over range ` ⌧ had ÊÊ ÊÊ »» »» ATLAS best fit ± 1 𝜎 < FIG. 7. Signal strenght measured by ATLAS in each of the four signal regions, as well as thepredictions from our two benchmark points, B1 and B2. The prediction from the best tan β fits ofB1 and B2 are shown as red and blue dots, respectively. Moreover, tan β is varied to encompassthe 1 σ range favored by data in the 2 ‘ τ had channel. Only values of tan β consistent with flavorconstraints are used, leading to a shorter allowed range for B2. Note that the signal strengthsdisplayed here are subject to up to a factor of 2 uncertainty, stemming from our MC estimation ofsignal efficiencies. flavor observables. For this reason, we cut off the range of B2 at this exclusion boundary.We can see that the typical excess pattern seen by [48] is well explained by the hypoth-esis of contamination from rare top decays. At this point, however, the uncertainties arestill large enough that even the no signal hypothesis is marginally consistent with observa-tions. Upcoming analyses with more data will either tighten the exclusions of our model, or,optimistically, corroborate the deviations from the SM expectation.17 I. DISCUSSION
While the LHC has so far found no compelling evidence for new physics, tremendouspossibilities still exist for discovery of new particles, including light electroweakly chargedones.In Type I two Higgs doublet models, a charged Higgs lighter than the top quark cannaturally be produced at significant rates from rare top decays. Remarkably, if there areadditional light neutral scalars in the spectrum, the final state for such signals has a substan-tial overlap with those of SM t ¯ tH SM processes. As a consequence, it is natural to considersignals from light extended Higgs sectors as a contaminant to existing SM searches.Interestingly, many - but not all - of the existing t ¯ tH SM searches show excesses, both atthe Tevatron as well as the 8 and 13 TeV runs of the LHC. It is challenging to simultaneouslyreconcile these excesses with each other - significant excesses in leptonic channels, and aninconclusive pattern in γγ and b ¯ b channels, for instance. If these excesses persist, theycould potentially be explained by the contamination of a new, charged Higgs signal from topdecays.On general grounds, one would expect that the more tuned a given analysis is to thespecific final state kinematics of the t ¯ tH SM process, the less sensitive it should be to BSM topdecays. That would be the case of LHC searches for t ¯ t ( H SM → γγ ), as these focus on a narrow m γγ window around 125 GeV; or of (post Higgs discovery) MVA analyses in general. On theother hand, more inclusive analyses should be more prone to contamination from signals ofextended Higgs sectors. Examples of more inclusive analyses are the Tevatron’s CDF searchfor t ¯ t ( h → b ¯ b ) [38, 39], which considers a broader m b ¯ b window of 100 −
150 GeV; and the13 TeV ATLAS search for leptonic t ¯ tH SM [48], which employs a more conventional cut-and-count strategy, and might have non-negligible acceptance to BSM signals in multileptonsplus two or more b -jets.We have recast the 13 TeV ATLAS search for t ¯ tH SM in multilepton final states [48], andfound that its signal regions can be naturally contaminated by a light Type I 2HDM spec-trum at low tan β . Our recast of the results of [48] provides new limits on these models.Furthermore, these models can also explain the excesses observed in the data without ex-18eeding null results from other measurements. In principle, this signal could also show up inother searches, such as those targeting H SM → b ¯ b , depending on the details of the MVA used,and the masses of A /H . Indeed, considering the high branching ratio of A /H → b ¯ b , finalstates with many b -jets may provide the strongest tests going forward.Should the excesses persist, it is clear that light Type I 2HDM spectra provide an attrac-tive potential explanation. Broadening search windows, especially for A /H → b ¯ b at lower m b ¯ b , could further constrain this scenario, or, possibly, provide the first evidence at the LHCof physics beyond the Standard Model. ACKNOWLEDGMENTS
We thank A. Djouadi and M. Spira for answering our questions regarding charged Higgs 3-body decays, and Kyle Cranmer for helpful discussions on statistics. We also thank A. Pierceand M. Graesser for interesting conversations about the ATLAS ttH excess. NW and DAare supported by the NSF under grants PHY-0947827 and PHY-1316753. SEH is supportedby the Cluster of Excellence Precision Physics, Fundamental Interactions and Structure ofMatter (PRISMA-EXC 1098) and by the Mainz Institute for Theoretical Physics.
Appendix A: Recasting the ATLAS search for t ¯ tH SM in multileptons1. Overview The production of a Higgs boson in association with two top quarks is expected in theSM and has therefore been searched for by ATLAS and CMS [42, 44, 47, 48]. In our study,we focus on the ATLAS 13 TeV search in multilepton final states [48], since it is based ona traditional cut-and-count analysis that is possible to recast. This search is targeted atevents where the Higgs decays to either
W W ∗ , ZZ ∗ , or τ + τ − . All these channels can leadto signatures with up to four leptons, for which the backgrounds are extremely low at theLHC. The potentially interesting events are grouped into four different signal regions: twowith a pair of same-sign light leptons and either zero or one hadronic tau (2 ‘ τ and 2 ‘ τ ),19ne with three light leptons (3 ‘ ) and one with four light leptons (4 ‘ ). The 13 . − datahave shown an excess in both the 2 ‘ τ and the 2 ‘ τ regions, more particularly for eventswith only one b -tagged jet. The peculiar structure of this excess cannot be explained solelyby an enhanced t ¯ tH SM cross section and could be a sign for new physics.Naively, the ( bW )( bW ∗ τ τ ) final state associated with our Type I 2HDMs is similar to the tt ( H SM → τ τ ) signal that is looked for in this search. The kinematics of the final states ishowever significantly different. Notably, only one W boson arises from the direct decay of atop quark. The other one is produced through the decay of the charged Higgs to H /A andis therefore much softer or even off-shell. Consequently, when the mass splitting between H + and H /A is small, events with 4 leptons will have lower efficiencies for trigger andpreselection cuts. Similarly, the efficiency of our signal in the 3 ‘ region should be lower thanthat of ttH SM . The observed rates for the 2 ‘ τ and 2 ‘ τ regions should however remainsignificant. Broadly speaking, this structure is similar to the excess observed in ATLAS.
2. The recasting procedure
In the ATLAS t ¯ tH SM search, events are selected and classified into four exclusive signalregions associated with various cuts on the multiplicity and momenta of leptons, hadronictaus, light jets and b -jets. In addition to these cuts, vetoes against m ‘‘ ’ m Z as well aslow mass leptonic Drell-Yan are imposed in events with same-flavor lepton pairs. Dependingon the signal region, the light flavor leptons can also be required to verify specific isolationcriteria. Although most of the selection cuts are straightforward, the tagging of hadronic tausas well as the isolation requirements on the light leptons require a more careful treatmentthat we detail in what follows.In the 2 ‘ τ , 2 ‘ τ and 3 ‘ regions, the light leptons have to pass loose and tight isolationcuts. Since these cuts are associated with efficiencies of 99% and 96% respectively [49],we do not implement them in our analysis. In the 4 ‘ region, the electrons and muons aresubmitted to so-called “gradient” isolation cuts, with a p T -dependent efficiency. These cutscan have a significant impact at low p T and we therefore take them into account when using Delphes [50] for detector simulation. 20 ‘ τ ‘ τ ‘ ‘N PGS N Delphes N ATLAS t ¯ t ( H SM → τ τ ) in the signal regions of [48], obtained using Delphes and
PGS for simulation of the detector response, as well as the expected yield provided byATLAS.
To estimate the errors associated with our modeling of the detector response, we use twodifferent detector simulators:
Delphes 3 and
PGS 4 [50, 51]. In
Delphes , we set the b - and τ -tagging efficiencies/mistag rates to the values used by ATLAS in [48] and initially loosenthe electron and muon isolation criteria. The rest of the parameters are set to the defaultvalues given in the ATLAS 13 TeV card from Delphes . In order to take into account thepossible loss of low p T electrons or muons in the 4 ‘ signal region, we select the light leptonsin this region with efficiencies corresponding to the ones of the gradient isolation cuts. Whengenerating events with PGS , we use the default lepton tagging algorithm and do not applyany additional isolation cuts. For hadronic taus, however, since the efficiency of the
PGS tagging algorithm is much lower than the one used by ATLAS, we modify the
PGS code toidentify all jets within ∆ R = 0 . ex post facto , given by the branching-ratio-weighted average of theone-prong and three-prong efficiencies given in [48]. Likewise, we modify the PGS b -taggingalgorithm to better represent the working point used in [48]. All the other cuts besides thepreselection cuts are implemented without any change.In order to validate our analysis, we generate t ¯ tH SM events using MadGraph5 [52], andstudy each of the Higgs decay modes ( τ + τ − , W W ∗ and ZZ ∗ ) independently. We match theseevents up to one additional jet and shower them with Pythia6 [53] using MLM matching [54]with the k T shower scheme [55]. As mentioned above, we use both PGS 4 [51] and
Delphes3 [50] to simulate the detector response. The expected event yields obtained by our MC studyfor each of the three Higgs decay channels are given in tables I, II and III. We observe that21 ‘ τ ‘ τ ‘ ‘N PGS N Delphes N ATLAS t ¯ t ( H SM → W W ∗ ) in the signal regions of [48], obtained using Delphes and
PGS for simulation of the detector response, as well as the expected yield provided byATLAS. 2 ‘ τ ‘ τ ‘ ‘N PGS N Delphes N ATLAS t ¯ t ( H SM → ZZ ∗ ) in the signal regions of [48], obtained using Delphes and
PGS for simulation of the detector response, as well as the expected yield provided byATLAS.
Delphes and
PGS exhibit complementary performances.
Delphes gives a better modeling ofthe ATLAS efficiencies in the 2 ‘ τ region, whereas PGS shows better agreement in the 2 ‘ τ ,3 ‘ and 4 ‘ regions. Overall, PGS provides a better modeling of the SM t ¯ tH SM efficiencies, andtherefore we use PGS for recasting the results in [48] in terms of rare top decays to b H + .In our analysis, we use the recasting procedure as is. That is, we do not apply any ad hoc correction factors to the efficiencies, since we could not determine exactly the origin of oursmall discrepancies. The overall uncertainties in our detector simulation translate into afactor of ∼
3. Signal generation
A charged Higgs can be produced along with a top quark and a bottom quark throughtwo distinct processes. First, as highlighted throughout this paper, this final state can22 g btH + FIG. 8. Feynman diagram for non-resonant production of H + ¯ tb . arise from on-shell top pair-production, with one of the tops decaying to H + b . As thecharged Higgs gets heavier than about 155 GeV, however, the branching ratio Br( t → b H + )gets suppressed, and off-shell production of the charged Higgs must be included as it givesan important contribution to the signal rate. The dominant non-resonant H ± productionprocess, pp → ¯ t b H + , is shown in Fig. 8. We include this process in our study for all chargedHiggs masses above 150 GeV, and use the NLO cross sections provided in [56].We use MadGraph [52] to generate t ¯ t + j events, and to decay both tops to (¯ bW − )( b H + )(and the corresponding charge conjugate process). We match these events up to one addi-tional jet and shower them using Pythia [53] with MLM matching [54] and the k T showerscheme [55]. We use Pythia to further decay the W ’s, H ± ’s and A ’s. For the detectorsimulation, we use PGS [51] with the same settings as described above. The non-resonantprocess tbH ± → (¯ bW − ) bH ± and its charge conjugate are generated with MadGraph . Thesubsequent steps are the same as for the first process.
Appendix B: Signal yields and statistical procedure1. Signal yields
As discussed in Appx. A, we use MC to obtain the signal efficiencies in each of the foursignal regions defined by ATLAS, for both the t ¯ t on-shell process: pp → ¯ t t → ¯ t b ( H + → W +( ∗ ) ( A → τ + τ − )) , (2.1)23s well as the non-resonant process in Fig. 8 pp → ¯ t b ( H + → W +( ∗ ) ( A → τ + τ − )) . (2.2)We generically denote the respective efficiencies by (cid:15) res and (cid:15) nonres .The signal yield for a specific signal region is then given by: S = S res + S nonres , (2.3)where S res = (cid:15) res × ( σ t ¯ t × L ) × (cid:16) t → b H + ) × Br( H + → W +( ∗ ) A ) × Br( A → τ + τ − ) (cid:17) , (2.4)and S nonres = (cid:15) nonres × σ nonres tan β × L ! × (cid:16) Br( H + → W +( ∗ ) A ) × Br( A → τ + τ − ) (cid:17) . (2.5)For the total cross sections, we use σ t ¯ t (cid:12)(cid:12)(cid:12)
13 TeV = 830 pb, and the following approximate fitfor σ nonres extracted from [56]: σ nonres ( m H ± ) (cid:12)(cid:12)(cid:12)(cid:12)
13 TeV = (cid:18) . − . m H ± GeV (cid:19) pb (2.6)in the range m H ± ∼ [150 −
2. Branching ratios
Throughout our study we set Br( A → τ + τ − ) = 0 . b H + is given by [5]:Br( t → b H + ) = R H + R H + , (2.7)where R H + = Γ( t → b H + )Γ( t → b W + )= p H + p W + β ( m t + m b − m H + )( m t + m b ) − m b m t m W ( m t + m b − m W ) + ( m t − m b ) . (2.8)24bove, p H + is the momentum of H + in the top’s rest frame, p H + = m t vuut − m H + m t ! − m b m t m H + m t ! + m b m t , (2.9)and p W + is defined in an analogous way.The branching ratios of the charged Higgs are computed from its various widths. Forcompleteness, we list all of them below.The charged Higgs width to a pair of light on-shell SM fermions is given by:Γ( H + → f ¯ f ) = G F √ π m H + tan β N c | U f ¯ f | (cid:16) m f + m f (cid:17) , (2.10)where N c = 3 and U f ¯ f is an element of the CKM matrix if f, ¯ f are quarks, and N c = 1 and U f ¯ f = 1 otherwise; m f , m ¯ f above are the running masses at µ = m H + .Since we are interested in spectra where m H + < m t , the width Γ( H + → t ∗ ¯ b ) goes via anoff-shell top quark, and is given by [22]:Γ( H + → W + b ¯ b ) = 3 G F m t π m H + tan β κ W κ t (4 κ W κ t + 3 κ t − κ W ) log κ W ( κ t − κ t − κ W ! + (3 κ t − κ t − κ W + 1) log (cid:18) κ t − κ t − κ W (cid:19) −
52 (2.11)+ 1 − κ W κ t (3 κ t − κ t κ W − κ t κ W + 4 κ W ) + κ W (4 − κ W / ! , where κ t = m t /m H + , and κ W = m W /m H + . Above, we neglect the contribution from thebottom yukawa coupling.The bosonic widths of the charged Higgs, Γ( H + → W +( ∗ ) φ ), depend on whether thefinal state W + is on- or off-shell. In the former case, we have:Γ( H + → W + φ ) = (1 − ζ φ ) G F √ π m W m H + q λ ( m φ , m W , m H + ) λ ( m φ , m H + , m W ) , (2.12)where ζ φ for φ = H SM , H , A are given in Sec. II, and λ ( x, y, z ) = (cid:18) − xz − yz (cid:19) − x yz . (2.13)On the other hand, if the W + is off-shell, we have [22]:Γ( H + → W + ∗ φ ) = (1 − ζ φ ) m H + G F m W π G m φ m H + , m W m H + ! , (2.14)25here G ( x, y ) = 18 − − x + y ) q λ G ( x, y ) π " y (1 + x − y ) − λ G ( x, y )(1 − x ) q λ G ( x, y ) + (cid:16) λ G ( x, y ) − x (cid:17) log( x ) + (1 − x )3 y (cid:16) y (1 + x ) − y + 2 λ G ( x, y ) (cid:17) , (2.15)and λ G ( x, y ) = − x + 2 y − ( x − y ) . (2.16)
3. Statistical treatment of fits and exclusions
For each signal region in [48] with N sr observed events, µ B sr ± ∆ B sr expected backgroundevents, and S sr expected signal events, we define the likelihood function: L sr ( S sr , B sr ) = P ( N sr | S sr + B sr ) × G ( B sr | µ B sr , ∆ B sr ) , (2.17)where P and G are Poisson and Gaussian distributions, respectively. The likelihood for thecombination of all four signal regions is given by the product of the individual likelihoods: L ( µ S , θ B ) = Y srj L srj ( S srj , B srj ) , (2.18)where θ B = ( B sr1 , B sr2 , B sr3 , B sr4 ), and µ S = ( m H + , m A , tan β ) uniquely specifies a pointin the model parameter space, and unambiguously determines the signal yield S srj in eachsignal region. Since the correlations between background uncertainties in the four signalregions were not provided in [48], we treat these uncertainties as uncorrelated. We note,however, that the post-fit backgrounds in [48] did not change substantially relative to theirpre-fit counterparts. This, combined with the factor of 2 uncertainties in our MC modelingof signal efficiencies, leads us to expect that the our results would not change substantiallywere we to properly include all background correlations in our fits.For obtaining limits, we use a profiled log-likelihood analysis. First, we define: λ ( µ S ) = log L ( µ S , ˆˆ θ B ) − log L (ˆ µ S , ˆ θ B ) , (2.19)where the unconditional likelihood estimators ˆ µ S , ˆ θ B maximize the global log L ( µ S , θ B ), andthe conditional estimator ˆˆ θ B maximizes log L ( µ S , θ B ) for a given µ S . Since there are 326ndependent degrees of freedom in µ S , namely, m H + , m A , and tan β , the p -value for aspecific model defined by µ S is determined by: p = 1 − CDF( χ , − λ ( µ S )) , (2.20)where CDF( χ , − λ ( µ S )) is the cumulative distribution function for a χ -distribution with 3degrees of freedom, evaluated at − λ ( µ S ). From the p -value we can determine the exclusionconfidence level for all points in the studied parameter space.Finally, we note that when finding the goodness of fit of a given mass point µ S =( m H + , m A ), we profile over the value of tan β that maximizes the log-likelihood, i.e., weuse λ ( µ S ) = log L ( µ S , tan ˆˆ β, ˆˆ θ B ) − log L (ˆ µ S , tan ˆ β, ˆ θ B ) . (2.21)In this case, we define the p -value in an analogous way, p = 1 − CDF( χ , − λ ( µ S )) , (2.22)but instead use a χ -distribution with only 2 degrees of freedom. [1] Further searches for squarks and gluinos in final states with jets and missing transverse momen-tum at √ s =13 TeV with the ATLAS detector , Tech. Rep. ATLAS-CONF-2016-078 (CERN,Geneva, 2016).[2] Search for electroweak SUSY production in multilepton final states in pp collisions atsqrt(s)=13 TeV with 12.9/fb , Tech. Rep. CMS-PAS-SUS-16-024 (CERN, Geneva, 2016).[3] T. Enomoto and R. Watanabe, JHEP , 002 (2016), arXiv:1511.05066 [hep-ph].[4] M. Misiak and M. Steinhauser, (2017), arXiv:1702.04571 [hep-ph].[5] J. F. Gunion, H. E. Haber, G. L. Kane, and S. Dawson, Front. Phys. , 1 (2000).[6] D. S. M. Alves, P. J. Fox, and N. J. Weiner, (2012), arXiv:1207.5499 [hep-ph].[7] M. Czakon and A. Mitov, Comput. Phys. Commun. , 2930 (2014), arXiv:1112.5675 [hep-ph].
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