The Hagedorn Temperature Revisited.

# The Hagedorn Temperature Revisited.

J. Cleymans    D. Worku
###### Abstract

The Hagedorn temperature, is determined from the number of hadronic resonances including all mesons and baryons. This leads to a stable result MeV consistent with the critical and the chemical freeze-out temperatures at zero chemical potential. We use this result to calculate the speed of sound and other thermodynamic quantities in the resonance hadron gas model for a wide range of baryon chemical potentials following the chemical freeze-out curve. We compare some of our results to those obtained previously in other papers.

25.75,-q.05.70,-a,64.70,-p,64.90,+b

## 1 Introduction

In 1965 Hagedorn   proposed that the number of hadronic resonances increases exponentially with the mass of the resonances. The idea, which was debated strongly when first proposed, has since been widely accepted and discussed by many authors . The concept was based on the assumption that the observed increase in the number of hadronic resonances would continue towards higher and higher masses as more experimental data became available . The scale of the exponential increase determines the value of the Hagedorn temperature, . Recent papers   have used the latest results from the Particle Data Group  to revisit the original analysis of Hagedorn to update the value of . This resulted in a surprising wide spread of possible values, with large variations as to whether one considers mesons or baryons with values ranging from MeV to MeV depending on the parametrization used and on the set of hadrons (mesons or baryons). There thus exists uncertainty as to the value of the Hagedorn temperature. These have two origins:

• the analytical form of the Hagedorn spectrum, especially the factor multiplying the exponential.

The first item will probably never be resolved satisfactorily due to the width of resonances and also due to their large number making it difficult to identify them. Splitting the spectrum into baryons and mesons further decreases the quality of the fits to the mass dependence of the mass spectrum. We therefore propose to stick to the original analysis of Hagedorn and consider a sum over all resonances, baryons, mesons, strange, non-strange, charm, bottom etc.. This is the state that is produced at the Large Hadron Collider (LHC), namely, a hadronic ensemble containing all possible resonances. The result is shown in Fig. 1 and leads to a good determination of because the range in is reasonably large extending up to 3 GeV before reaching a plateau presumably due to the parsity of hadronic resonances above this value. Details about the parametrization used will be presented below.

The Hagedorn temperature naturally leads to the notion that hadronic matter cannot exceed a limiting temperature and increasing the beam energy in protonproton and protonantiprotonÌ collisions results in more and more hadronic resonances being produced without a corresponding increase in the temperature of the final (freeze-out) state. This is the situation observed at the highest energies at the LHC. It was suggested long ago  that the Hagedorn temperature is connected to the existence of a different phase in which quarks and gluons are no longer confined. At present, the Hagedorn temperature is often understood as the temperature of the phase transition from hadrons to a quark-gluon plasma.
A recent analysis  investigated the hadron resonance gas model  to show the stability of various thermodynamic quantities in heavy-ion collisions when the Hagedorn spectrum is used literally, i.e. without a cut-off on the number of resonances beyond a certain mass. In particular, the authors explored the addition of the hadron resonance gas model including an exponentially large number of undiscovered resonances which are naturally included in the Hagedorn model. Their results showed that the hadron resonance gas model gave different results for thermodynamic quantities but the overall chemical analysis was reasonably stable.

Furthermore, the use of so-called Hagedorn States (HS), based on the exponentially increasing spectrum, close to the critical temperature can explain fast chemical equilibration by HS regeneration  and provide a unique method to compare lattice results for using thermal fits HS provide a lower than thermal fits without HS . These authors estimated effects by extending the hadron mass spectrum beyond GeV for = 200 MeV and assume that high mass excited states produce one kaon while producing multiple pions thus further reducing the .

In this paper, we extend previous work  by extending explicit expressions of relevant thermodynamic quantities for non-zero chemical potentials. In particular the speed of sound, , can be considered as a sensitive indicator of the critical behavior in strongly interacting matter. The results show a sharp dip of in the critical region, which is an indication that thermodynamics in the vicinity of confinement is indeed driven by the higher excited hadronic states.
The outline of this paper is as follows. In section 2, we present the influence of the Hagedorn spectrum on the hadron yields to find the thermodynamic parameters and explain the basic concepts used in this paper. In section 3, we derive the number, energy and entropy densities and also the speed of sound. In section 4, we show the results using the hadron and its extension resonance gas model for particular thermodynamic quantities and discuss the relationship of with temperature and chemical potentials and compare thermodynamic quantities, like energy and entropy density. The last section covers the conclusion.

## 2 Motivation

The particle data table contains hundreds of hadronic resonances  including the well known stable hadrons like nucleons, pions, hyperons (, , ), kaons etc.. In our calculations we used a list of hadronic resonances including in total baryons and antibaryons, mesons and antimesons (counted without considering their isospin and spin degeneracies). Most of the hadronic resonances decay quickly via strong interactions before reaching the detector, hence they are usually identified via their decay products. The mass of a decaying particle is equal to the total energy of the products measured in its rest frame.
The basic idea of this paper is to add resonances using the Hagedorn model for the spectrum. Using the hadronic data we show the relationship between the number of hadronic resonances and the mass in Fig. 1 where we took hadrons with masses up to GeV. The density of states obtained this way has been fitted using the following equation ;

 ρh(m)=c(m2+m20)5/4exp(mTH), (1)

where and are constant parameters given in the table below

## 3 The Hadron Resonance Gas Model and its Extension

The thermodynamic properties of the Hadron Resonance Gas Model (HRGM) can be determined from the partition function

 lnZ(V,T,μ) = ∫dm[ρM(m)lnZb(m,V,T,μ) (2) +ρB(m)lnZf(m,V,T,μ)],

where the gas is contained in a volume , has a temperature and chemical potential , and are the partition functions for an ideal gas of bosons and fermions respectively with mass , and are the spectral density of mesons and baryons. By using Eq. 2, one can compute the number denisty , energy density , entropy density , pressure , speed of sound and specific heat .
hadron properties enter these models through . The HRGM model takes the observed spectrum of hadrons up to some cutoff of mass , defined by

 ρM,B(m)=mi≤λ∑igiδ(m−mi), (3)

where are the masses of the known hadrons and the degeneracy factor. In order to explore the stability of results obtained using the HRGM, variant of these models is often used in which one takes the observed spectrum of states up to a certain cutoff of mass and above this one includes an exponentially rising cumulative density of hadron states, which is defined in Eq. 2. This defines the Extended Hadron Resonance Gas Model (EHRGM). The density of states becomes

 ρh(m)=mi≤λ∑igiδ(m−mi)+c(m2+m20)5/4exp(mTH)θ(m−λ), (4)

where the model parameter and are fitted to data on the cumulative distribution of the sets of hadrons, . Typically this model uses GeV. Moreover, the parameters are determined from the hadronic spectrum with masses up to GeV as shown in Fig. 1. The results for and are given in Table 1.

## 4 Derivation of thermodynamic quantities in EHRGM

We are considering here the Boltzmann distribution for simplicity in order to present some of the results in a compact form. The partition function for a single particle is given by

 lnZ=gVTm22π2K2(mT)exp(μT), (5)

where is modified Bessel function. The meson mass distribution is taken to be given by

 ρM,B(m)=mi<λ∑i=1gM,Biδ(m−mi)+c(m2+m20)5/4exp(mTH)θ(m−λ), (6)

### 4.1 Derivation for the particle densities and pressure

The particle density can be written as the sum of the two terms

 n=nM+nB, (7)

where and are number densities of mesons and baryons respectively. They are defined by

 nM,B = T2π2mi<λ∑i=1exp(μiT)[gM,Bim2iK2(miT)+ (8) c∫∞m=λm2(m2+m20)5/4exp(mTH)K2(mT)dm],

where , for our case we consider an isospin symmetric system, where GeV. The pressure is given by

 P=T∂lnZM∂V+T∂lnZB∂V=PM+PB, (9)

where and are the pressure of mesons and baryons respectively.

### 4.2 Derivation for energy density

The energy density is given by

 εM,B=T2V∂lnZM,B∂T, (10)

where is the modified Bessel function, and are energy densities of mesons and baryons respectively.

 (11)

where

 AH=c∫∞mm2(m2+m20)5/4exp(mTH)K2(mT)dm.

From the above expressions, we can obtain the entropy density

 s=ε+P−nSμS−nBμBT, (12)

where and are the net number densities for strange and baryonic particles respectively

 nS(B)=TV∂lnZ∂μS(B). (13)

### 4.3 Derivation of the speed of sound

In hydrodynamic models the speed of sound plays an important role in the evolution of a system and is an ingredient in the understanding of the effects of a phase transition . It is well-known  that the speed of sound has to be calculated at constant entropy per particle . This makes the calculation more complicated than the one at zero chemical potential where it is sufficient to keep the temperature fixed. In our extension we take the condition  into account for non-zero baryon and strangeness chemical potentials, imposing overall strangeness zero. The squared speed is thus calculated starting from

 C2s(T,μ)=(∂P∂ε)s/n, (14)

The complete expression for the speed of sound can be rewritten as

 C2s(T,μ)=(∂P∂T)+(∂P∂μS)(dμSdT)+(∂P∂μB)(dμBdT)(∂ε∂T)+(∂ε∂μS)(dμSdT)+(∂ε∂μB)(dμBdT), (15)

where the derivatives and can be evaluated from two conditions. The first condition comes from the requirement that the ratio must be kept fixed,

 sn=0, (16)

while the second condition follows from the conservation of strangeness

 nS=n¯S. (17)

The resulting expressions for and are presented in detail in the appendix.

## 5 Results using HRGM and EHRGM

There is not much difference between the HRGM and EHRGM models at low temperature, since the heavy resonances do not play an important role there. Alternatively, at high temperatures, , it should agree with the results of the lattice simulations of QCD . In our case, we have determined the temperature and chemical potential dependencies of the pressure and the number, energy and entropy densities using the chemical freeze-out curve . Using these results, we have then calculated the speed of sound. The thermodynamic variables obtained using HRGM, we plot the energy and entropy densities scaled by the appropriate powers of , which is shown in Fig. 2a,  3a and  4a, at the GeV didn’t observe any sudden change of the thermodynamic variables and close to the critical temperature, it showed smooth shape as the temperature goes beyong .
Moreover, when we use the EHRGM, we see a different behaviour at the critical temperature which is shown in Fig. 2b,  3b and  4b. These show that if the system undergoes a first-order phase transition, both temperature and pressure remain constant as hadronic matter is converted from hadron gas to quark-gluon plasma; the energy and entropy density, however, change discontinuously. This leads to show us a new state of matter.

The value of the speed of sound remains well below the ideal-gas limit for massless particles , even at very high temperatures, energy densities and pressure. In Fig. 4b represents, when it crosses the transition point in the case of full QCD, we expect as before that should vanish beside that is inversely proportional to the specific heat , and this can lead to a divergence at the critical temperature . Of course, due to finite volume effects, the velocity of sound will most likely not completely vanishes at .

The velocity of sound versus the temperature as shown in Fig. 4b is to evaluate , we have followed and first expressed and in physical units, using EHRGM in Eq. 4. The temperature and chemical potential dependence of all the relevant thermodynamic quantities shows unique behaviour at the critical point , specially for the speed of sound, there is a pronounced dip as evidence for the phase transition in the system. Based on our calculations, the velocity of sound at is different from zero; it is expected to become zero but due to finite volume effects.

## 6 Conclusions

In this paper we have made a new analysis of the number of hadronic resonances using the latest information from the Particle Data Group . This leads to a temperature which is consistent with the most recent results based on lattice QCD estimates of the phase transition temperature  and also the chemical freeze-out temperature at zero baryon density . We have extended calculations of the speed of sound to non-zero baryon and strangeness chemical potentials keeping fixed . This is done for both the hadronic resonance gas model (HRGM) and the extended hadronic resonance gas model (EHRGM) which includes the exponentially increasing spectrum of hadrons following the Hagedorn parametrization. The EHRGM shows that the speed of sound goes to zero at the phase transition point while the HRGM shows a smooth dip followed by an increase.

## Appendix A Appendix: Speed of sound at non-zero chemical potentials.

The speed of sound is given by

 C2s(T,μ)=(∂P∂ε)sn, (A.1)

where is the entropy per particle which is kept fixed. using the variables and , this can be rewritten as

 C2s(T,μ)=(∂P∂T)+(∂P∂μS)(dμSdT)+(∂P∂μB)(dμBdT)(∂ε∂T)+(∂ε∂μS)(dμSdT)+(∂ε∂μB)(dμBdT), (A.2)

where the derivatives and can be evaluated using two conditions. The first condition comes from keeping the ratio constant. From the derivative one obtains

 d(sn)=0, (A.3)

which implies

 nds=sdn. (A.4)

In terms of , and this equation can be written as

 n(∂s∂T)dT + n(∂s∂μB)dμB+n(∂s∂μS)dμS (A.5) = s(∂n∂T)dT+s(∂n∂μB)dμB+s(∂n∂μS)dμS.

divide the above expression by on both sides, so that it becomes

 n(∂s∂T) + n(∂s∂μB)(dμBdT)+n(∂s∂μS)(dμSdT) (A.6) = s(∂n∂T)+s(∂n∂μB)(dμBdT)+s(∂n∂μS)(dμSdT).

Rearranging the above expression in order to write in terms of one obtains

 [n(∂s∂μB)−s(∂n∂μB)](dμBdT) =s(∂n∂T)−n(∂s∂T)−[n(∂s∂μS)−s(∂n∂μS)](dμSdT). (A.7)

Defining

 A = n(∂s∂T)−s(∂n∂T), (A.8) B = n(∂s∂μB)−s(∂n∂μB), (A.9) C = n(∂s∂μS)−s(∂n∂μS). (A.10)

The final expression for condition one is

 dμBdT=−1B[A+CdμSdT]. (A.11)

The second condition comes from overall strangeness neutrality, which is

 nS=n¯S (A.12)

where and are the strange and antistrange particle densities. Similarly the derivative of equation Eq. A.12 should thus satisfy

 d(nS)=d(n¯S) (A.13)

this implies equation Eq. A.13 can be expressed as

 (∂nS∂T)+(∂nS∂μB)(dμBdT)+(∂nS∂μS)(dμSdT) =(∂n¯S∂T)+(∂n¯S∂μB)(dμBdT)+(∂n¯S∂μS)(dμSdT). (A.14)

We can apply the same method as above to write in terms of for the above relation

 [(∂nS∂μB)−(∂n¯S∂μB)](dμBdT) (A.15) = (∂n¯S∂T)−(∂nS∂T)−[(∂nS∂μS)−(∂n¯S∂μS)](dμSdT),

we define , hence it represents that the number of strangeness density for baryons and mesons

 L=nBS+nMS,

and , the number of antistrangeness density for baryons and mesons

 R=n¯BS+n¯MS.

Define now

 E = ∂L∂μB−∂R∂μB, (A.16) F = ∂L∂μS−∂R∂μS, (A.17) D = ∂L∂T−∂R∂T. (A.18)

Hence, the final expression for condition two becomes

 dμBdT=−1E[D+FdμSdT]. (A.19)

Finally, by equating equation  A.11 and  A.19 we find

 dμSdT=AE−BDBF−CE, (A.20)

and

 dμBdT=CD−AFBF−CE, (A.21)

Which is the relation used in the text.

## References

• 1. R. Hagedorn, Supplemento al Nuovo Cimento Volume III, 147 (1965); Nuovo Cimento 35 (1965) 395; Nuovo Cimento 56 A (1968) 1027.
• 2. R. Hagedorn and J. Ranft, Nucl. Phys. B 48 (1972) 157-190.
• 3. N. Cabibbo and G. Parisi, Phys. Lett. B 59 (1975) 67.
• 4. G. Veneziano, Nuovo Cimento 57A (1968) 190.
• 5. K. Bardakci and S. Mandelstam, Phys. Rep.184 (1969) 1640;
S. Fubini and G. Veneziano, Nuovo Cimento 64A (1969) 811.
• 6. K. Huang and S. Weinberg, Phys. Rev. Lett. 25 (1970) 895.
• 7. H. Satz, Phys. Rev. D 19 (1979) 1912.
• 8. Ph. Blanchard, S. Fortunato and H. Satz, Eur. Phys. J. C 34 (2004) 361.
• 9. R. V. Gavai and A. Goksch, Phys. Rev. D 33 (1986) 614.
• 10. K. Redlich and H. Satz, Phys. Rev. D 33 (1986) 3747.
• 11. F. Karsch, E. Laermann and A. Peikert, Phys. Lett. B 478 (2000) 447.
• 12. Particle Data Group, C. Caso et al., Eur. Phys. J. C 3, (1998) 1-794
• 13. M. Chojnacki, W. Florkowski and T. Csorgo, Phys. Rev. C 71 (2006) 044902.
• 14. M. Chojnacki and W. Florkowski Acta Physica Polonica B 38 (2007) 3249.
• 15. S. Chatterjee, S. Gupta and R. M. Godbole, Phys. Rev. C 81 044907 (2010).
• 16. W. Broniowski, W. Florkowski and L. Y. Glozman, Phys. Rev. D 70 (2004) 117503
• 17. W. Broniowski, Preprint hep-ph 0008112, in “Bled 2000: Few Quark Problems”.
• 18. W. Broniowski and W. Florkowski, Phys. Lett. B 490 (2000) 223-227
• 19. J. Cleymans, H. Oeschler, K. Redlich, S. Wheaton, Phys. Rev. C 73 (2006) 034905.
• 20. A. Andronic, P. Braun-Munzinger, J. Stachel, Nucl. Phys. A 772 (2006) 167.
• 21. F. Becattini, J. Manninen, M. Gazdzicki, Phys. Rev. C 73 (2006) 044905.
• 22. J. Noronha-Hostler, M. Beitel, C. Greiner, I. Shovkovy, Phys. Rev. C 81 (2010) 054909.
• 23. J. Noronha-Hostler, H. Ahmad, J. Noronha, C. Greiner, Phys. Rev. C 82 (2010) 024913.
• 24. A. Andronic, P. Braun-Munzinger, J. Stachel, Phys. Lett. B 673 (2009) 142
• 25. P. Castorina, K. Redlich and H. Satz, Europ. Phys. J. C 59 (2009) 67.
• 26. J. Noronha-Hostler, J. Noronha, C. Greiner, Phys. Rev. Lett. 103 (2009) 172302.
• 27. P. Castorina, J. Cleymans, D.E. Miller, H. Satz, Eur. Phys. J. C 66 (2010) 207-213.
e-Print: arXiv:0906.2289 [hep-ph]
• 28. J. Cleymans, N. Bilic, E. Suhonen and D.W. von Oertzen, Phys. Lett. B 311 (1993) 266-272.
• 29. L.D. Landau, E.M. Lifshitz (1987). Fluid Mechanics. Vol. 6 (2nd ed.). Butterworth-Heinemann. ISBN 978-0-080-33933-7.
• 30. S. Borsanyi et al. Wuppertal-Budapest Collaboration, JHEP 1011 (2010) 077;
• 31. W. Söldner (for the HotQCD collaboration) [arXiv:1012.4484] [hep-lat]].
• 32. P. Braun-Munzinger and J. Stachel, Nucl. Phys. A 606 (1996) 320; D. Prorok and L. Turko, arXiv:hep-ph/0101220.
You are adding the first comment!
How to quickly get a good reply:
• Give credit where it’s due by listing out the positive aspects of a paper before getting into which changes should be made.
• Be specific in your critique, and provide supporting evidence with appropriate references to substantiate general statements.
• Your comment should inspire ideas to flow and help the author improves the paper.

The better we are at sharing our knowledge with each other, the faster we move forward.
The feedback must be of minimum 40 characters and the title a minimum of 5 characters