---
abstract: 'The energy spectrum of nucleons in high-density nuclear matter is investigated in the framework of relativistic meson-nucleon many-body theory, employing the $1/N$ expansion method.  The coupling of the nucleon with the particle-hole excitations in the medium flattens the spectra in the vicinity of the Fermi surface.  The effect grows logarithmically for increasing density and eventually leads to instability of the normal state.  The validity of the mean-field theory at high density is criticized.' 
author:
- |
    Kazuhiro TANAKA[^1]\
    \
    [*Radiation Laboratory*]{}\
    [*The Institute of Physical and Chemical Research (RIKEN)*]{}\
    [*Hirosawa 2-1, Wako-shi, Saitama, 351-01 Japan*]{}
date:  
title: 'Quasiparticle properties and the dynamics of high-density nuclear matter'
---

3.7ex

One universal character shared by most normal Fermi liquids is the enhancement of the effective mass around the Fermi surface [@ma].  Typical indications observed empirically are the density of single-particle levels around the “Fermi surface” in nuclei, and the logarithmic ($T^{3}\ln T$) correction to the specific heat at low temperature $T$ in normal liquid $^{3}$He.  It is now well established that these phenomena due to the energy-dependence of the effective mass reflect the dynamics beyond the mean-field theory: The relevant mechanism is the coupling of the single-particle motion to other excitation modes, e.g., particle-hole (ph) excitations in the case of nuclear matter.  As a result, the nucleon effective mass at normal density is close to the free mass $M$ or slightly larger at the Fermi surface while it is about the traditional value ($\sim 0.7 M$) away from the Fermi surface. 

The effective mass at the Fermi surface determines the density of states at the Fermi surface.  Therefore, the dynamical couplings to the single-particle motion could influence the observables significantly.  In this light, the density-dependence, especially the high-density behavior, of the effective mass would be relevant to descriptions of high-density matter, needed to understand, e.g., stellar structure or evolution and high-energy heavy ion reactions. 

The investigation of the high-density behavior of the effective mass requires a relativistic framework of Fermi liquids.  In this paper we compute the quasiparticle spectra including higher-order many-body correlations employing a relativistic many-body theory: In order to exhibit clearly the points avoiding complicated aspects such as the compositeness of hadrons and the possibilities of new degrees of freedom at high density, we use the simplest relativistic model of nuclear matter, the “${{\sigma}}{{\omega}}$ model” [@se]: $${\cal L} = \overline{\psi} \left( i {{\mbox{$\!\not\!\partial$}}}
- g_{{{\omega}}} {{\mbox{$\!\not\!{{\omega}}$}}} - M + g_{{{\sigma}}} {{\sigma}}\right) \psi
+ \frac{1}{2} \left( \partial_{\mu} {{\sigma}}\right)^{2}
- \frac{1}{2} m_{{{\sigma}}}^{2} {{\sigma}}^{2} - \frac{1}{4}{{\omega}}_{\mu \nu}^{2}
+ \frac{1}{2} m_{{{\omega}}}^{2} {{\omega}}_{\mu}^{2},
\label{eq:la}$$ with ${{\omega}}_{\mu \nu} = \partial_{\mu}{{\omega}}_{\nu} - \partial_{\nu} {{\omega}}_{\mu}$. $\psi$, ${{\sigma}}$ and ${{\omega}}_{\mu}$ are the nucleon, ${{\sigma}}$- and ${{\omega}}$-meson fields. 

Since this model is renormalizable, it allows a systematic investigation of the many-body correlations.  One convenient scheme for this purpose is the $1/N$ expansion where $N$ is the extended number of nucleon species [@ta3].  The expansion can be obtained by applying the rescaling: $g_{{{\sigma}}} \rightarrow 1/\sqrt{N} g_{{{\sigma}}}$, $g_{{{\omega}}} \rightarrow 1/\sqrt{N} g_{{{\omega}}}$; ${{\sigma}}\rightarrow \sqrt{N} {{\sigma}}$, ${{\omega}}_{\mu} \rightarrow \sqrt{N} {{\omega}}_{\mu}$ in the lagrangian [(\[eq:la\])]{}.  The energy density of nuclear matter up to the next-to-leading order is given by ${{\cal E}}= {{\cal E}}_{0} + {{\cal E}}_{\rm e} + {{\cal E}}_{\rm c}$ [@ta3; @ta4].  ${{\cal E}}_{0}$ denotes the Hartree term which is the leading order of $O(N)$, while ${{\cal E}}_{\rm e}$ and ${{\cal E}}_{\rm c}$ denote the exchange and the RPA-type correlation energy terms, both of which are of $O(1)$. The formulae for these terms can be found in ref. [ @ta3].  Based on this framework, a successful description of the saturation property, the equation of state, and the Fermi-liquid properties of nuclear matter has been obtained [@ta3; @ta4; @he], improving the Hartree description.  The essential ingredients are the many-body correlation effects due to ${{\cal E}}_{\rm c}$.

The quasiparticle energy corresponding to our energy density can be obtained by taking the functional derivative with respect to the Fermi distribution function $n({\mbox{\boldmath {$p$}}})$ following the Landau theory of Fermi liquids [@la1] and its relativistic extension [@ba3]: $$\epsilon ({\mbox{\boldmath {$p$}}}) = \frac{\delta {{\cal E}}}{\delta n({\mbox{\boldmath {$p$}}})}
= \epsilon_{0} ({\mbox{\boldmath {$p$}}}) + \epsilon_{\rm e}({\mbox{\boldmath {$p$}}})
+ \epsilon_{\rm c}({\mbox{\boldmath {$p$}}}),
\label{eq:qp2}$$ where the three terms derive from the corresponding three terms of our energy density.  The familiar Hartree term $\epsilon_{0}$, which is of $O(1)$, contains the scalar and the vector mean-field potential contributions.  The other contributions of $O(1/N)$ are given by [@ta4] $$\epsilon_{\rm x}({\mbox{\boldmath {$p$}}}) = \frac{{{M^{\ast}}}}{E^{\ast}(p)}
\overline{u}(p) {{\mit \Sigma}}_{\rm x}(p) u(p)
\;\;\; \;\;\; ({\rm x} = {\rm e}, \; {\rm c}),
\label{eq:e1}$$ where $E^{\ast}(p) = \sqrt{{\mbox{\boldmath {$p$}}}^{2} + {M^{\ast}}^{2}}$ and ${{M^{\ast}}}= M - g_{{{\sigma}}} \overline{{{\sigma}}}$ with $\overline{{{\sigma}}}$ the ground-state expectation value of the ${{\sigma}}$-field.  $u(p)$ is the Hartree spinor normalized by $\overline{u}u = 1$. ${{\mit \Sigma}}_{\rm x}$ is the nucleon self-energy shown in fig.  \[fig:fig2\]; ${\rm x} = {\rm e}$ and ${\rm c}$ correspond to the first (exchange) and the following (RPA) graphs, respectively.[^2]

The counter terms to renormalize the divergences of the self-energies are determined through eq. [( \[eq:qp2\])]{} consistently with the renormalization conditions for the energy density: We renormalize all vertex parts in free space at the scales $q^{2} = 0$ and ${{\mbox{$\!\not\!p$}}} = M$ for the meson and the nucleon legs, respectively.  The Landau ghost singularity appearing in the one-loop meson propagators is removed by the method based on the Källén-Lehmann representation proposed by Redmond [@re] and extended to finite density in refs. [ @ta1; @ta3; @ta4]. 

If the nucleon Hartree propagator connecting the external legs in the diagrams of fig.  \[fig:fig2\] is separated into the “Feynman part” (F) and the “density-dependent part” (D) [@se], we obtain the conventional separation of ${{\mit \Sigma}}_{{{\rm x}}}$ into ${{\mit \Sigma}}_{{{\rm x}}{{\rm F}}}$ and ${{\mit \Sigma}}_{{{\rm x}}{{\rm D}}}$. For the present purpose, it is convenient to employ a new separation of the self-energy based on a Wick rotation (WR) of the loop integration in ${{\mit \Sigma}}_{{{\rm x}}{{\rm F}}}$ [@ta4; @bl]: $\epsilon_{{{\rm x}}}({\mbox{\boldmath {$p$}}}) = \epsilon_{{{\rm x}}{{\rm W}}}({\mbox{\boldmath {$p$}}})
+ \epsilon_{{{\rm x}}{{\rm P}}}({\mbox{\boldmath {$p$}}})$, where $\epsilon_{{{\rm x}}{{\rm W}}}$ involves the Wick rotated euclidian 4-integral, and $\epsilon_{{{\rm x}}{{\rm P}}}$ contains an on-shell intermediate nucleon line, arising from the nucleon pole due to the WR and the contribution of ${{\mit \Sigma}}_{{{\rm x}}{{\rm D}}}$. $\epsilon_{{{\rm x}}{{\rm P}}}$ can be expressed compactly: $$\epsilon_{{{\rm x}}{{\rm P}}} ({\mbox{\boldmath {$p$}}})= 4 \int^{{p_{{{\rm F}}}}}_{p}
\frac{{{\rm d}}q q^{2}}{(2 \pi)^{2}} \int^{1}_{-1}
{{\rm d}}\!  \cos \theta \: f_{{{\rm x}}}({\mbox{\boldmath {$p$}}}, {\mbox{\boldmath {$q$}}}),
\label{eq:po}$$ with $\theta$ the angle between ${\mbox{\boldmath {$p$}}}$ and ${\mbox{\boldmath {$q$}}}$. $f_{{{\rm x}}}({\mbox{\boldmath {$p$}}}, {\mbox{\boldmath {$q$}}})$ are the exchange-type interactions between two quasiparticles with momenta ${\mbox{\boldmath {$p$}}}$, ${\mbox{\boldmath {$q$}}}$. $f_{\rm e}$ involves the noninteracting meson propagators, while $f_{\rm c}$ takes into account the modifications due to the one-loop meson self-energies. 

The slope of the quasiparticle spectra at the Fermi surface defines the Fermi velocity: $${v_{{{\rm F}}}}= \left.  \frac{\partial \epsilon({\mbox{\boldmath {$p$}}})}{\partial |{\mbox{\boldmath {$p$}}}|}
\right|_{|{\mbox{\boldmath {$p$}}}| = {p_{{{\rm F}}}}}
= \frac{{p_{{{\rm F}}}}}{E^{\ast}({p_{{{\rm F}}}})} + v_{{{\rm F}}\rm e} + v_{{{\rm F}}\rm c},
\label{eq:vf}$$ where the three terms derive from the corresponding terms of eq. [( \[eq:qp2\])]{}.  In the NR treatment, ${v_{{{\rm F}}}}\equiv {p_{{{\rm F}}}}/M_{\rm eff}$ with $M_{\rm eff}$ the effective mass at the Fermi surface.  If, as we will show below, the enhancement of $M_{\rm eff}$ becomes stronger and stronger with increasing density, $v_{\rm F}$ can eventually become a decreasing function of the density.  In particular, $v_{\rm F} \rightarrow 0$ means that the spectra become flat around the Fermi surface. 

In fig.  \[fig:fig4\] we show ${v_{{{\rm F}}}}$ as a function of ${p_{{{\rm F}}}}$. The parameters are taken from refs. [ @ta3; @ta4] and are shown in table \[tab:tabn\]: The dotted line shows the Hartree approximation, i.e., the first term of eq. [( \[eq:vf\])]{} using the “Hartree” parameter set; it approaches the causal limit “1” at high density.  In contrast, the solid line, which shows the full result of [(\[eq:vf\])]{} using the parameter set B, crosses ${v_{{{\rm F}}}}= 0$ at the critical ${p_{{{\rm F}}}}= p_{{{\rm F}}}^{{{\rm c}}}$, and ${v_{{{\rm F}}}}$ becomes negative for higher densities.  The dashed line, including only one density-dependent bubble instead of the full RPA propagator,[^3] shows a behavior similar to the full case, but $p_{{{\rm F}}}^{{{\rm c}}}$ becomes smaller.  We also show the full result of [(\[eq:vf\])]{} using the parameter set A by the dot-dashed line; the comparison with the solid line demonstrates the sensitivity of $p_{{{\rm F}}}^{{{\rm c}}}$ to the parameters (see eq. [( \[eq:d3\])]{} below).  In fig.  \[fig:fig5\] we show the quasiparticle energy of eq. [( \[eq:qp2\])]{} for the case of the solid line of fig.  \[fig:fig4\].  For high densities, the spectra show the “anomaly” around the Fermi surface, i.e., they are flat or decreasing due to the inclusion of the higher-order self-energy terms.[^4] Thus, the effect grows with increasing density. 

From fig.  \[fig:fig5\](b), we see that this effect is due to $\epsilon_{{{\rm c}}{{\rm P}}}$, which is a decreasing function around the Fermi surface. ( This is seen already at normal density, but for higher densities this effect becomes more pronounced and shows up in the total results.)  The decrease of $\epsilon_{{{\rm c}}{{\rm P}}}$ around the Fermi surface could be expected from eq. [( \[eq:po\])]{} if the result of the angular average of $f_{{{\rm x}}} ({\mbox{\boldmath {$p$}}}, {\mbox{\boldmath {$q$}}})$ assumes a positive value (see also ref. [ @bl] for the NR case). 

Now can we understand the origin of the growth of the anomaly due to $\epsilon_{{{\rm c}}{{\rm P}}}$?  For this purpose it will be useful to examine the leading high-density behavior of ${v_{{{\rm F}}}}$. We set for ${p_{{{\rm F}}}}\rightarrow \infty$ ${v_{{{\rm F}}}}\rightarrow 1 + \delta v_{{{\rm F}}{{\rm e}}}
+ \delta v_{{{\rm F}}{{\rm c}}} \equiv 1 + \delta v_{{{\rm F}}}$, where “1” is due to the first (Hartree) term of eq. [( \[eq:vf\])]{} while $\delta v_{{{\rm F}}{{\rm x}}}$ (${{\rm x}}= {{\rm e}}$, ${{\rm c}}$) denote the asymptotic limits of the following two ($O(1/N)$) terms $v_{{{\rm F}}{{\rm e}}}$, $v_{{{\rm F}}{{\rm c}}}$: $$\delta v_{{{\rm F}}{{\rm x}}} \approx
- \frac{p_{{{\rm F}}}^{2}}{\pi^{2}}
\int_{-1}^{1} {{\rm d}}\!  \cos \theta \: \left. 
f_{{{\rm x}}} ({\mbox{\boldmath {$p$}}}, {\mbox{\boldmath {$q$}}})
\right|_{p = q = {p_{{{\rm F}}}}}. 
\label{eq:vfi}$$ The r.h.s.  is the result obtained by substituting $\epsilon_{{{\rm x}}{{\rm P}}}$ of eq. [( \[eq:po\])]{} into eq. [( \[eq:vf\])]{}. “ $\approx$” means that the contribution due to $\epsilon_{{{\rm x}}{{\rm W}}}$ is neglected because of the above discussion of fig.  \[fig:fig5\]. ( In fact, it can be shown analytically that $\epsilon_{{{\rm x}}{{\rm W}}}$ does not alter our conclusions [@ta5].) 

It is convenient to separate the integrand of eq. [( \[eq:vfi\])]{} according to the types of the possible meson modes in the medium [@ta4]: $$\left.  f_{{{\rm x}}}({\mbox{\boldmath {$p$}}}, {\mbox{\boldmath {$q$}}}) \right|_{p = q = {p_{{{\rm F}}}}}
= - \frac{1}{8 p_{{{\rm F}}}^{2}}
\sum_{{{\rm I}}} A_{{{\rm x}}{{\rm I}}}(\cos \theta)
h_{{{\rm I}}}(\cos \theta),
\label{eq:fs}$$ where ${{\rm I}}= {{\sigma}}, {{\rm M}}, {{\rm L}}$ and ${{\rm T}}$, corresponding to the ${{\sigma}}$-, mixed, longitudinal and transverse contributions.  $h_{{{\rm I}}}$ is a dimensionless function accounting for the scalar and the vector vertex structures.  $A_{{{\rm e}}{{\rm I}}}$ are the exchange contributions due to the noninteracting ${{\sigma}}$- and ${{\omega}}$-meson propagators while $A_{{{\rm c}}{{\rm I}}}$ give the modifications of $A_{{{\rm e}}{{\rm I}}}$ due to the RPA correlations. 

To extract the leading asymptotic behavior of [(\[eq:fs\])]{}, we set ${{M^{\ast}}}= 0$; because there is no singularity for ${{M^{\ast}}}= 0$ and from the dimension counting, any contributions omitted here will be suppressed for ${p_{{{\rm F}}}}\rightarrow \infty$ compared to the retained ones.  In this limit it can be easily seen that the contributions due to the M-mode are suppressed compared to the others. 

In this case, the relevant contributions (${{\rm I}}= {{\sigma}}$, ${{\rm L}}$ and ${{\rm T}}$) have the form: $$a^{n}_{{{\rm I}}} = \frac{g_{{{\rm I}}}^{2}}{2}
\frac{1}{t + \kappa_{{{\rm I}}}^{2}}
\left(\frac{\hat{{{\mit \Pi}}}_{{{\rm I}}}(t)}{t + \kappa_{{{\rm I}}}^{2}}
\right)^{n},
\label{eq:yu}$$ where $t = 1- \cos \theta$, $g_{{{\rm L}}} = g_{{{\rm T}}} \equiv g_{{{\omega}}}$, $m_{{{\rm L}}}= m_{{{\rm T}}} \equiv m_{{{\omega}}}$, $\kappa_{{{\rm I}}}^{2} = m_{{{\rm I}}}^{2}/2 p_{{{\rm F}}}^{2}$, and $\hat{{{\mit \Pi}}}_{{{\rm I}}}(t) = \left. {{ \mit \Pi}}_{{{\rm I}}}(p-q)
\right|_{p = q = {p_{{{\rm F}}}}}/2 p_{{{\rm F}}}^{2}$ with ${{\mit \Pi}}_{{{\rm I}}}$ the 1-loop self-energies for those modes.[^5] $a_{{{\rm I}}}^{n}$ give the contributions involving $n$ bubbles for the ${{\rm I}}$-mode, such that $A_{{{\rm e}}{{\rm I}}}= a_{{{\rm I}}}^{0}$ and $A_{{{\rm c}}{{\rm I}}} = \sum_{n \ge 1} (-1)^{n - 1} a_{{{\rm I}}}^{n}$.

If the integrand of eq. [( \[eq:vfi\])]{} given by eqs. [( \[eq:fs\])]{}, [(\[eq:yu\])]{} were finite everywhere in the integration region, the result would coincide with the naive dimension counting which leads to $\delta v_{{{\rm F}}{{\rm x}}} \rightarrow {\rm const}$ as $p_{\rm F} \rightarrow \infty$. For ${p_{{{\rm F}}}}\rightarrow \infty$, however, $\kappa_{{{\rm I}}}^{2} \rightarrow 0$ and therefore the integral diverges at $\theta = 0$. This divergence implies that $\delta v_{{{\rm F}}{{\rm x}}}$ could contain positive powers or logarithms of ${p_{{{\rm F}}}}$. (It can be shown [@ta5] that there is no other singularity.)  $\kappa_{{{\rm I}}}^{2} \rightarrow 0$ corresponds formally to zero meson masses; therefore, these singularities for $\theta \simeq 0$, which now prove to give the most dominant contributions for ${p_{{{\rm F}}}}\rightarrow \infty$, correspond to the infrared singularity in the massless limit. 

By substituting the analytic formulae of $h_{{{\rm I}}}$ [@ta4] and ${{\mit \Pi}}_{{{\rm I}}}$ [@ta1; @ku] into eqs. [( \[eq:vfi\])]{}-[(\[eq:yu\])]{} and by extracting the singularities for $\theta \simeq 0$, we can obtain $\delta v_{{{\rm F}}{{\rm x}}}$ up to logarithmic accuracy.  For ${{\rm x}}= {{\rm e}}$, we obtain $$\begin{aligned}
\delta v_{{{\rm F}}{{\rm e}}} &=& \left(g_{{{\sigma}}}^{2}\times O(1)\right)
+ \left(\frac{1}{2}\left(\frac{g_{{{\omega}}}}{2 \pi}\right)^{2}
\left\{\ln\frac{p_{{{\rm F}}}^{2}}{m_{{{\omega}}}^{2}} + O(1) \right\}\right)
\nonumber \\
&+& \left(- \frac{1}{2} \left(\frac{g_{{{\omega}}}}{2 \pi}\right)^{2}
\left\{\ln\frac{p_{{{\rm F}}}^{2}}{m_{{{\omega}}}^{2}} + O(1) \right\}\right). 
\label{eq:d1}\end{aligned}$$ The three terms show the contributions due to the relevant three modes ${{\sigma}}$, ${{\rm L}}$ and ${{\rm T}}$, respectively.  Though both the contributions due to the L- and T-modes grow logarithmically, reflecting the logarithmic infrared divergence at $\theta \rightarrow 0$, they cancel out, leading to $\delta v_{{{\rm F}}{{\rm e}}} = {\rm const}$. The contribution to $\delta v_{{{\rm F}}{{\rm c}}}$ due to one bubble insertion, i.e., due to the $n = 1$ term of eq. [( \[eq:yu\])]{}, can be obtained similarly: $$\begin{aligned}
\delta v_{{{\rm F}}1 {\rm b}} &=&
\left(g_{{{\sigma}}}^{4} \times O(\ln p_{{{\rm F}}}^{2})\right)
+ \left(
- \left( \frac{g_{{{\omega}}}}{2 \pi}\right)^{4}
\left\{ \frac{4 p_{{{\rm F}}}^{2}}{m_{{{\omega}}}^{2}} - \frac{1}{3}
\ln \frac{p_{{{\rm F}}}^{2}}{M^{2}}
\ln\frac{p_{{{\rm F}}}^{2}}{m_{{{\omega}}}^{2}}+ O(\ln p_{{{\rm F}}}^{2})
\right\}\right)
\nonumber \\
&+& \left(- \left(\frac{g_{{{\omega}}}}{2 \pi}\right)^{4}
\left\{ \frac{1}{3} \ln \frac{p_{{{\rm F}}}^{2}}{M^{2}}
\ln \frac{p_{{{\rm F}}}^{2}}{m_{{{\omega}}}^{2}} + O(\ln p_{{{\rm F}}}^{2})
\right\}
\right). 
\label{eq:d2}\end{aligned}$$ The coupling to the ph excitations makes the infrared divergences stronger;[^6] the most dominant contribution is due to the linearly divergent integral due to the L-mode, and the result grows like $p_{{{\rm F}}}^{2}$. However, it can be easily seen that the insertion of more bubbles causes stronger infrared divergences, leading to terms of higher powers in ${p_{{{\rm F}}}}$.

This situation is reminiscent of the high-density NR electron gas [@ge].  This suggests that we have to sum up all the bubble graphs, forming the complete RPA series.  Therefore, [*the infrared structure of $v_{{{\rm F}}}$ due to the coupling of the ph excitations naturally leads to our $1/N$ expansion scheme.*]{}  By summing up all the bubbles, we obtain for the total $\delta v_{{{\rm F}}}$: $$\begin{aligned}
\delta v_{{{\rm F}}} &=& \left(g_{{{\sigma}}}(p_{{{\rm F}}})^{2}\times O(1)\right)
+ \left(g_{{{\omega}}}(p_{{{\rm F}}})^{2} \times O(1) \right)
\nonumber \\
&+& \left(- \frac{1}{2} \left(\frac{g_{{{\omega}}}(p_{{{\rm F}}})}{2 \pi}\right)^{2}
\left\{\ln\frac{p_{{{\rm F}}}^{2}}{m_{{{\omega}}}^{2}} + O(1) \right\}\right). 
\label{eq:d3}\end{aligned}$$ The resummation of the ring graphs renders the infrared divergence mild, leaving only the logarithmically growing term due to the ${{\rm T}}$-mode (compare the dashed and the solid curves of fig.  \[fig:fig4\]).  Thus the decrease of $v_{{{\rm F}}}$, which led to the inversion of the spectrum of fig.  \[fig:fig5\], is due to the logarithmic correction [(\[eq:d3\])]{} due to the self-energies beyond the mean-field theory. 

The resummation also replaces the original coupling constants $g_{{{\rm I}}}$ renormalized at the meson momenta $q^{2} = 0$ by the running coupling constants $g_{{{\rm I}}}({p_{{{\rm F}}}})$ at the scale ${p_{{{\rm F}}}}$ [@ta5]: Before applying Redmond’s method to avoid the Landau ghost [@re], they are related by the one-loop formulae: for ${p_{{{\rm F}}}}\gg M$, $g_{{{\omega}}}({p_{{{\rm F}}}})^{2} \simeq g_{{{\omega}}}^{2} \left/
\left(1 - \frac{g_{{{\omega}}}^{2}}
{6 \pi^{2}} \ln\frac{p_{{{\rm F}}}^{2}}{M^{2}} \right) \right. ,$ and similarly for $g_{{{\sigma}}}({p_{{{\rm F}}}})$. $g_{{{\rm I}}}({p_{{{\rm F}}}})$ of eq. [( \[eq:d3\])]{} should be understood as the result of Redmond’s method; they grow for increasing ${p_{{{\rm F}}}}$ toward finite values for the bare coupling constants [@re].  Thus the growth of the last term of eq. [( \[eq:d3\])]{} is enhanced by the growth of $g_{{{\omega}}}(p_{{{\rm F}}})$.

Our result can be summarized by the behavior of the screening mass in the medium.  The infrared behavior of the ${{\rm I}}$-mode contribution is governed by the screening mass (squared) $m_{{{\rm sc}}{{\rm I}}}^{2}$, which is obtained from the self-energy ${{\mit \Pi}}_{{{\rm I}}}(q)$ by letting $q^{0} \rightarrow 0$ first followed by ${\mbox{\boldmath {$q$}}}\rightarrow 0$ (see eqs. [( \[eq:vfi\])]{}-[(\[eq:yu\])]{}).  The Lorentz invariance guarantees $m_{{{\rm sc}}{{\rm L}}}^{2} = m_{{{\rm sc}}{{\rm T}}}^{2}$, which would lead to the cancellation between the ${{\rm L}}$- and ${{\rm T}}$-mode contributions (see eq. [( \[eq:d1\])]{}).  However, the coupling of the ph excitations breaks the invariance: The ${{\rm L}}$-mode acquires a nonzero Debye screening mass, which is $m_{{{\rm sc}}{{\rm L}}}^{2} = 2 g_{{{\omega}}}^{2} p_{{{\rm F}}}^{2}/\pi^{2}$ for ${p_{{{\rm F}}}}\rightarrow \infty$, but $m_{{{\rm sc}}{{\rm T}}}^{2} = 0$ due to baryon current conservation [@be].  As a result, the cancellation between the two modes becomes incomplete, leaving the result of eq. [( \[eq:d3\])]{}. 

Now we discuss physical interpretations of our results. 

Above the critical $p_{{{\rm F}}}^{{{\rm c}}}$ where $v_{\rm F} = 0$, the normal state is unstable.  In our calculation of the ground-state energy density and the quasiparticle energy we assumed the normal state [@la1].  If $v_{{{\rm F}}} < 0$, however, this assumption is not valid.  This suggests a stability condition: ${v_{{{\rm F}}}}> 0$ for a normal Fermi liquid, in addition to the wellknown conditions for the dimensionless Landau-Migdal parameters [@la1].  If the ground state is not normal, one should re-compute the energy density self-consistently allowing the possibility of new configurations like the so-called “Fermi-gap” state: If one literally accepts the results of fig.  \[fig:fig5\], this indicates a 1st order phase transition at ${p_{{{\rm F}}}}= p_{{{\rm F}}}^{{{\rm c}}}$ from the normal state to the Fermi-gap state.[^7] In contrast to the Fermi sphere for the normal state, the nucleons occupy the momentum states within the inner sphere and the outer shell for the Fermi-gap state.  In our model, $p_{{{\rm F}}}^{{{\rm c}}} = 4.07$fm$^{-1}$ which would imply the critical density for neutron matter $\varrho^{{{\rm c}}} {\stackrel{\displaystyle{>}}{\sim}}15{\varrho}_{0}$.

We also note another possible scenario: Near the critical point $v_{\rm F} = 0$, the terms of higher order than we considered here, which would give contributions $\sim ( g_{{{\omega}}}(p_{{{\rm F}}})^{2} \ln p_{{{\rm F}}}^{2} )^{n}$ ($n \ge 2$), also could be important.  It would be a very hard job to compute these contributions.  Intuitively, one could regard the result of eq. [( \[eq:d3\])]{} as the leading approximation of ${v_{{{\rm F}}}}\approx 1/
\left(1 + \left( g_{{{\omega}}}({p_{{{\rm F}}}})/2 \pi\right)^{2}/2
\times \ln \left(p_{{{\rm F}}}^{2}/m_{{{\omega}}}^{2}\right) \right)$. In this case ${v_{{{\rm F}}}}= 0$ would not occur for any finite ${p_{{{\rm F}}}}$; ${v_{{{\rm F}}}}\rightarrow 0$ asymptotically. 

In connection with these scenarios, we stress that the (formal) high-density limit based on the lagrangian [(\[eq:la\])]{} does not correspond to the mean-field theory, contrary to the usual claim [@se]: For the first scenario this is obvious; the second one implies an infinite density of states at the Fermi surface for ${p_{{{\rm F}}}}\rightarrow \infty$, which never be reached in the mean-field theory.  The usual claim tacitly assumes the absence of infrared singularities, which reflect large long-wavelength fluctuations of the fields, from the higher order corrections.  This assumption leads to the validity of the naive dimension counting, and therefore to the dominance of the mean-field contributions because they contain the largest number of independent fermion loops in each order of the interaction.  As discussed above, the effects which lead to the modification of this naive argument are due to the strong infrared behavior induced in the high-density medium. 

This conclusion for the high-density behavior seems to be inevitable for relativistic descriptions: The relevant infrared behavior is due to the ph excitation in the vector channel (see eqs. [( \[eq:d1\])]{}-[(\[eq:d3\])]{}).[^8] However, the vector boson degrees of freedom are indispensable to keep matter from collapsing at high density by supplying the repulsion.  We also note that the conclusion might be valid even if the compositeness of hadrons were taken into account: Even though the vertex form factors were introduced, they could not influence the results dominated by the contribution from the infrared region ($q^{2} = 0$). 

Though we have shown the possibility of the Fermi-gap state, the transition density $\varrho^{{{\rm c}}}$ appears to be above the critical densities to more familiar new phases like quark matter phase, pion condensed phase, etc.  However, our result was obtained just within a simple model, and $\varrho^{{{\rm c}}}$ is very sensitive even to small corrections, which might be due to higher orders in $1/N$ or other hadronic degrees of freedom like $\pi$, $\rho$, etc.  After the inclusion of these effects, the competition or coexistence of the Fermi-gap state with more familiar new phases would be an interesting aspect of high-density hadronic matter. 

In conclusion, we have investigated dynamical effects in high-density nuclear matter due to the higher order many-body correlations.  The anomaly in the fermion spectra, corresponding to the famous effective mass enhancement in NR Fermi liquids, grows with increasing density.  The effect is due to the strong infrared behavior caused by the coupling of the ph excitation to the single-particle motion.  The picture of high-density matter emerged from our investigation is much more complicated than the mean-field description.  Though a more realistic treatment would be required to fix the relevance to phenomenology, the universal character of this phenomenon already poses various interesting aspects for high-density matter, like the possibility of the Fermi-gap state, the thermodynamic properties and transport phenomena in this phase, and the extension to gauge theories.  These points will be reported in future publications [@ta5]. 

The author expresses his thank to Dr. W. Bentz for helpful discussions.  He is also grateful to Prof. K. Yazaki and the members of nuclear theory group in RIKEN for valuable comments.  The work was performed under the auspices of Special Researchers’ Basic Science Program of RIKEN. 

[99]{}

C. Mahaux, P. F. Bortignon, R. A. Broglia and C. H. Dasso, Phys.  Reports [**120**]{} (1985) 1, and references therein. 

B. D. Serot and J. D. Walecka, Adv.  in Nucl.  Phys. [** 16**]{}, ed.  J. W. Negele and E. Vogt (Plenum, 1986) p. 1. 

K. Tanaka and W. Bentz, Nucl.  Phys. [** A540**]{} (1992) 383. 

K. Tanaka, W. Bentz and A. Arima, Nucl.  Phys. [** A555**]{} (1993) 151. 

P. J. Redmond, [ Phys.  Rev. ]{} [**112**]{} (1958) 1404;\
N. N. Bogoliubov, A. A. Logunov and D. V. Shirkov, Sov.  Phys.  JETP (Engl.  Trasl.) [** 37**]{} (1960) 574. 

K. Tanaka, W. Bentz, A. Arima and F. Beck, [ Nucl.  Phys. ]{} [** A528**]{} (1991) 676. 

W. Bentz, L. G. Liu and A. Arima, Ann.  of Phys.  (N.Y.) [**188**]{} (1988) 61. 

G. Hejc, H. Baier and W. Bentz, Phys.  Lett.  B, in press. 

L. D. Landau, Sov.  Phys.  JETP [**3**]{} (1956) 920; [**5**]{} (1957) 101; [**8**]{} (1959) 70;\
A. B. Migdal, Theory of finite Fermi systems and applications to atomic nuclei (Wiley, New York, 1967). 

G. Baym and S. A. Chin, Nucl.  Phys. [** A262**]{} (1976) 527. 

J. P. Blaizot and B. L. Friman, Nucl.  Phys. [** A372**]{} (1981) 69. 

H. Kurasawa and T. Suzuki, Nucl.  Phys. [** A445**]{} (1985) 685. 

M. Gell-Mann and K. A. Brueckner, Phys.  Rev. [**106**]{} (1957) 364. 

C. J. Horowitz and B. D. Serot, [ Phys.  Lett.  B ]{} [**109**]{} (1982) 341. 

K. Tanaka, in preparation. 

Figure captions {#figure-captions .unnumbered}
===============

1.   [Feynman graphs of the nucleon self-energy of $O(1/N)$. The dashed lines in the figure denote the noninteracting ${{\sigma}}$- and ${{\omega}}$-meson propagators; the solid lines denote the nucleon Hartree propagators on the background meson fields and the Fermi sea.]{}  \[fig:fig2\]

2.   [The Fermi velocity as a function of the Fermi momentum.  For explanation of the curves, see text.]{}  \[fig:fig4\]

3.   [The real parts of the quasiparticle energy (eq. [( \[eq:qp2\])]{}) are shown as functions of momentum by the solid lines. ( a), (b) and (c) are for cases of $p_{{{\rm F}}}=1.30$, $4.07$ and $4.50{\rm fm}^{-1}$, which respectively corresponds to ${v_{{{\rm F}}}}> 0$, $= 0$ and $<0$. The dot-dashed line shows the Hartree contribution contained in the full result.  The insertion in (b) shows the $O(1/N)$-contributions as functions of momentum.  The dot-dashed line shows $\epsilon_{{{\rm e}}}$ of eq. [( \[eq:qp2\])]{}, while the solid line shows $\epsilon_{{{\rm c}}}$. The dotted and dashed lines are the separated contributions $\epsilon_{{{\rm c}}{{\rm W}}}$ and $\epsilon_{{{\rm c}}{{\rm P}}}$ contained in the solid line.]{}  \[fig:fig5\]

Table captions {#table-captions .unnumbered}
==============

1.   [The parameters used for numerical computation.]{}  \[tab:tabn\]

Table 1 {#table-1 .unnumbered}
=======

              $\displaystyle{g_{\sigma}^{2}}/4 \pi$   $\displaystyle{ g_{\omega}^{2}}/4 \pi$  $m_{\sigma}$ [[\[MeV\]]{}]{}   $m_{\omega}$ [[\[MeV\]]{}]{}
  --------- --------------------------------------- ---------------------------------------- ------------------------------ ------------------------------
  Hartree                                      6.23                                     8.18              550                            783
                                                                                                                            
  $1/N$ A                                      2.24                                     3.65              550                            783
  B                                            2.60                                     4.22              650                            783
                                                                                                                            

[^1]: Special Researcher, Basic Science Program

[^2]: For convenience of the following discussion, we employ a slightly different grouping of the next-to-leading terms compared to the preceding works [@ta3; @ta4]. 

[^3]: The Feynman bubbles are summed up in this result. 

[^4]: The latter case with the negative ${v_{{{\rm F}}}}$ does not correspond to the normal state and therefore, strictly speaking, the Landau theory is not applicable.  This point will be discussed later. 

[^5]: In this work ${{\mit \Pi}}_{{{\rm L}}}$ denotes the minus of “${{\mit \Pi}}_{{{\rm L}}}$” in refs. [ @ta3; @ta4]. 

[^6]: “h” includes the negative-energy states. 

[^7]: The possibility of the Fermi-gap state for high-density nuclear matter was pointed out in ref. [ @ho] in the Hartree-Fock approximation of the lagrangian [(\[eq:la\])]{}.  However, as stressed above, the mechanism in our case is beyond the Hartree-Fock approximation (recall $\delta v_{{{\rm F}}{{\rm e}}} = {\rm const}$). 

[^8]: In view of this, the extension of our approach to gauge theories would be appealing, and is under investigation for cases of cold QED and QCD plasma with finite fermion density [@ta5].